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Instructions to authors Aims and scope Physics Reports keeps the active physicist up-to-date on developments in a wide range of topics by publishing timely reviews which are more extensive than just literature surveys but normally less than a full monograph. Each Report deals with one speciﬁc subject. These reviews are specialist in nature but contain enough introductory material to make the main points intelligible to a non-specialist. The reader will not only be able to distinguish important developments and trends but will also ﬁnd a suﬃcient number of references to the original literature. Submission In principle, papers are written and submitted on the invitation of one of the Editors, although the Editors would be glad to receive suggestions. Proposals for review articles (approximately 500–1000 words) should be sent by the authors to one of the Editors listed below. The Editor will evaluate proposals on the basis of timeliness and relevance and inform the authors as soon as possible. All submitted papers are subject to a refereeing process. Editors J.V. ALLABY (Experimental high-energy physics), EP Division, CERN, CH-1211 Geneva 23, Switzerland. E-mail: [email protected] D.D. AWSCHALOM (Experimental condensed matter physics), Department of Physics, University of California, Santa Barbara, CA 93106, USA. E-mail: [email protected] J.A. BAGGER (High-energy physics), Department of Physics & Astronomy, The Johns Hopkins University, 3400 North Charles Street, Baltimore MD 21218, USA. E-mail: [email protected] C.W.J. BEENAKKER (Mesoscopic physics), Instituut–Lorentz, Universiteit Leiden, P.O. Box 9506, 2300 RA Leiden, The Netherlands. E-mail: [email protected] G.E. BROWN (Nuclear physics), Institute for Theoretical Physics, State University of New York at Stony Brook, Stony Brook, NY 11974, USA. E-mail: [email protected] D.K. CAMPBELL (Non-linear dynamics), Dean, College of Engineering, Boston University, 44 Cummington Street, Boston, MA 02215, USA. E-mail: [email protected] G. COMSA (Surfaces and thin ﬁlms), Institut fur . Physikalische und Theoretische Chemie, Universit.at Bonn, Wegelerstrasse 12, D-53115 Bonn, Germany. E-mail: [email protected] J. EICHLER (Atomic and molecular physics), Hahn-Meitner-Institut Berlin, Abteilung Theoretische Physik, Glienicker Strasse 100, 14109 Berlin, Germany. E-mail: [email protected] M.P. KAMIONKOWSKI (Astrophysics), Theoretical Astrophysics 130-33, California Institute of Technology, 1200 East California Blvd., Pasadena, CA 91125, USA. E-mail: [email protected] M.L. KLEIN (Soft condensed matter physics), Department of Chemistry, University of Pennsylvania, Philadelphia, PA 19104-6323, USA. E-mail: [email protected]

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A.A. MARADUDIN (Condensed matter physics), Department of Physics and Astronomy, University of California, Irvine, CA 92697-4575, USA. E-mail: [email protected] D.L. MILLS (Condensed matter physics), Department of Physics and Astronomy, University of California, Irvine, CA 92697-4575, USA. E-mail: [email protected] H. ORLAND (Statistical physics and ﬁeld theory), Service de Physique Theorique, CE-Saclay, CEA, 91191 Gif-sur-Yvette Cedex, France. E-mail: [email protected] R. PETRONZIO (High-energy physics), Dipartimento di Fisica, Universita" di Roma – Tor Vergata, Via della Ricerca Scientiﬁca, 1, I-00133 Rome, Italy. E-mail: [email protected] S. PEYERIMHOFF (Molecular physics), Institute of Physical and Theoretical Chemistry, Wegelerstrasse 12, D-53115 Bonn, Germany. E-mail: [email protected] I. PROCACCIA (Statistical mechanics), Department of Chemical Physics, Weizmann Institute of Science, Rehovot 76100, Israel. E-mail: [email protected] E. SACKMANN (Biological physics), Physik-Department E22 (Biophysics Lab.), Technische Universit.at Munchen, . D-85747 Garching, Germany. E-mail: [email protected] A. SCHWIMMER (High-energy physics), Department of Physics of Complex Systems, Weizmann Institute of Science, Rehovot 76100, Israel. E-mail: [email protected] R.N. SUDAN (Plasma physics), Laboratory of Plasma Studies, Cornell University, 369 Upson Hall, Ithaca, NY 14853-7501, USA. E-mail: [email protected] W. WEISE (Physics of hadrons and nuclei), Institut fur . Theoretische Physik, Physik Department, Technische Universit.at Munchen, . James Franck Strae, D-85748 Garching, Germany. E-mail: [email protected] Manuscript style guidelines Papers should be written in correct English. Authors with insuﬃcient command of the English language should seek linguistic advice. Manuscripts should be typed on one side of the paper, with double line spacing and a wide margin. The character size should be suﬃciently large that all subscripts and superscripts in mathematical expressions are clearly legible. Please note that manuscripts should be accompanied by separate sheets containing: the title, authors’ names and addresses, abstract, PACS codes and keywords, a table of contents, and a list of ﬁgure captions and tables. – Address: The name, complete postal address, e-mail address, telephone and fax number of the corresponding author should be indicated on the manuscript. – Abstract: A short informative abstract not exceeding approximately 150 words is required. – PACS codes/keywords: Please supply one or more PACS-1999 classiﬁcation codes and up to 4 keywords of your own choice for indexing purposes. PACS is available online from our homepage (http://www.elsevier.com/locate/physrep). References. The list of references may be organized according to the number system or the nameyear (Harvard) system. Number system: [1] M.J. Ablowitz, D.J. Kaup, A.C. Newell and H. Segur, The inverse scattering transform – Fourier analysis for nonlinear problems, Studies in Applied Mathematics 53 (1974) 249–315. [2] M. Abramowitz and I. Stegun, Handbook of Mathematical Functions (Dover, New York, 1965).

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[3] B. Ziegler, in: New Vistas in Electro-nuclear Physics, eds E.L. Tomusiak, H.S. Kaplan and E.T. Dressler (Plenum, New York, 1986) p. 293. A reference should not contain more than one article. Harvard system:

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Physics Reports 381 (2003) 1 – 233 www.elsevier.com/locate/physrep

Hidden local symmetry at loop A new perspective of composite gauge boson and chiral phase transition Masayasu Harada∗ , Koichi Yamawaki Department of Physics, Nagoya University, Nagoya 464-8602, Japan Accepted 21 February 2003 editor: G:E: Brown

Abstract We develop an e/ective 0eld theory of QCD and QCD-like theories beyond the Standard Model, based on the hidden local symmetry (HLS) model for the pseudoscalar mesons () as Nambu–Goldstone bosons and the vector mesons () as gauge bosons. The presence of gauge symmetry of HLS is vital to the systematic low energy expansion or the chiral perturbation theory (ChPT) with loops of as well as . We 0rst formulate the ChPT with HLS in details and further include quadratic divergences which are crucial to the chiral phase transition. Detailed calculations of the one-loop renormalization-group equation of the parameters of the HLS model are given, based on which we show the phase diagram of the full parameter space. The bare parameters (de0ned at cuto/ ) of the HLS model are determined by the matching (“Wilsonian matching”) with the underlying QCD at through the operator-product expansion of current correlators. Amazingly, the Wilsonian matching provides the e/ective 0eld theory with the otherwise unknown information of the underlying QCD such as the explicit Nc dependence and predicts low energy phenomenology for the three->avored QCD in remarkable agreement with the experiments. Furthermore, when the chiral symmetry restoration takes place in the underlying QCD, the Wilsonian matching uniquely leads to the Vector Manifestation (VM) as a new pattern of Wigner realization of chiral symmetry, with the becoming degenerate with the massless as the chiral partner. In the VM the vector dominance is badly violated. The VM is in fact realized in the large Nf QCD when Nf → Nfcrit −0, with the chiral symmetry restoration point Nfcrit 5Nc =3 being in rough agreement with the lattice simulation for Nc = 3. The large Nf QCD near the critical point provides a concrete example of a strong coupling gauge theory that generates a theory of weakly coupled light composite gauge bosons. Similarly to the Seiberg duality in the SUSY QCD, the SU(Nf ) HLS plays a role of a “magnetic theory” dual to the SU(Nc ) QCD as an “electric theory”. The proof of the low energy theorem of the HLS at any loop order is intact even including quadratic divergences. The VM can be realized also in hot and/or dense QCD. c 2003 Elsevier Science B.V. All rights reserved. PACS: 12.38.−t; 12.60.−i ∗

Corresponding author. E-mail address: [email protected] (M. Harada).

c 2003 Elsevier Science B.V. All rights reserved. 0370-1573/03/$ - see front matter doi:10.1016/S0370-1573(03)00139-X

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

Contents 1. Introduction . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2. A brief review of the chiral perturbation theory . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.1. Generating functional of QCD . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2. Derivative expansion . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.3. Order counting . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.4. Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.5. Renormalization . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.6. Values of low energy constants . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.7. Particle assignment . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.8. Example 1: vector form factors and L9 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.9. Example 2: → e and L10 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3. Hidden local symmetry . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1. Necessity for vector mesons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2. Gglobal × Hlocal model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3. Lagrangian with lowest derivatives . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.4. Particle assignment . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.5. Physical predictions at tree level . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.6. Vector meson saturation of the low energy constants (relation to the ChPT) . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.7. Relation to other models of vector mesons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.7.1. Matter 0eld method . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.7.2. Massive Yang–Mills method . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.7.3. Anti-symmetric tensor 0eld method . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.8. Anomalous processes . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4. Chiral perturbation theory with HLS . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.1. Derivative expansion in the HLS model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.2. O(p2 ) Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.3. O(p4 ) Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.4. Background 0eld gauge . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5. Quadratic divergences . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5.1. Role of quadratic divergences in the phase transition . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5.2. Chiral restoration in the nonlinear chiral Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5.3. Quadratic divergence in symmetry preserving regularization . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.6. Two-point functions at one loop . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.7. Low-energy theorem at one loop . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.8. Renormalization group equations in the Wilsonian sense . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.9. Matching HLS with ChPT . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.10. Phase structure of the HLS . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5. Wilsonian matching . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.1. Matching HLS with the underlying QCD . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.2. Determination of the bare parameters of the HLS Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.3. Results of the Wilsonian matching . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.3.1. Full analysis . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.3.2. “Phenomenology” with a() = 1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.4. Predictions for QCD with Nf = 2 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.5. Spectral function sum rules . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6. Vector manifestation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.1. Vector manifestation (VM) of chiral symmetry restoration . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.1.1. Formulation of the VM . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.1.2. VM vs. GL (Ginzburg–Landau/Gell–Mann–Levy) manifestation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

3 11 11 13 13 15 18 19 19 20 22 23 24 24 27 28 30 36 38 38 45 48 50 58 59 62 63 68 73 74 83 86 87 94 96 98 102 108 109 112 119 119 124 127 130 135 136 136 140

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233 6.1.3. Conformal phase transition . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.1.4. Vector manifestation vs. “vector realization” . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.1.5. Vector manifestation only as a limit . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.2. Chiral phase transition in large Nf QCD . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.3. Chiral restoration and VM in the e/ective 0eld theory of large Nf QCD . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.3.1. Chiral restoration . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.3.2. Critical behaviors . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.3.3. Nf -dependence of the parameters for 3 6 Nf ¡ Nfcrit . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.3.4. Vector dominance in large Nf QCD . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.4. Seiberg-type duality . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7. Renormalization at any loop order and the low energy theorem . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.1. BRS transformation and proposition . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.2. Proof of the proposition . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.3. Low energy theorem of the HLS . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8. Towards hot and/or dense matter calculation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.1. Hadronic thermal e/ects . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.2. Vector manifestation at non-zero temperature . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.3. Application to dense matter calculation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 9. Summary and discussions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Acknowledgements . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix A. Convenient formulae . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.1. Formulae for Feynman integrals . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.2. Formulae for parameter integrals . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.3. Formulae for generators . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.4. Incomplete gamma function . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.5. Polarization tensors at non-zero temperature . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.6. Functions used at non-zero temperature . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix B. Feynman rules in the background 0eld gauge . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . B.1. Propagators . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . B.2. Three-point vertices . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . B.3. Four-point vertices . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix C. Feynman rules in the Landau gauge . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . C.1. Propagators . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . C.2. Two-point vertices (mixing terms) . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . C.3. Three-point vertices . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . C.4. Four-point vertices . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix D. Renormalization in the heat kernel expansion . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . D.1. Ghost contributions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . D.2. , V and contributions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . References . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

3 144 148 151 153 154 154 157 161 165 167 168 169 171 174 175 176 180 184 188 196 196 196 198 200 201 202 203 203 204 204 206 207 208 208 209 210 213 213 215 229

1. Introduction As is well known, the vector mesons are the very physical objects that the non-Abelian gauge theory was 0rst applied to in the history [203,165]. Before the advent of QCD the notion of “massive gauge bosons” was in fact very successful in the vector meson phenomenology [165]. Nevertheless, little attention was paid to the idea that the vector meson are literally gauge bosons, partly because of their non-vanishing mass. It is rather ironical that the idea of the vector mesons being gauge bosons

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was forgotten for long time, even after the Higgs mechanism was established for the electroweak gauge theory. Actually it was long considered that the vector meson mass cannot be formulated as the spontaneously generated gauge boson mass via Higgs mechanism in a way consistent with the gauge symmetry and the chiral symmetry. It was only in 1984 that hidden local symmetry (HLS) was proposed by collaborations including one of the present authors (K.Y.) [21–23,74] to describe the vector mesons as genuine gauge bosons with the mass being generated via Higgs mechanism in the framework of the nonlinear chiral Lagrangian. The approach is based on the general observation (see Ref. [24]) that the nonlinear sigma model on the manifold G=H is gauge equivalent to another model having a larger symmetry Gglobal × Hlocal , Hlocal being the HLS whose gauge 0elds are auxiliary 0elds and can be eliminated when the kinetic terms are ignored. As usual in the gauge theories, the HLS Hlocal is broken by the gauge-0xing which then breaks also the Gglobal . As a result, in the absence of the kinetic term of the HLS gauge bosons we get back precisely the original nonlinear sigma model based on G=H , with G being a residual global symmetry under combined transformation of Hlocal and Gglobal and H the diagonal sum of these two. In the case at hand, the relevant nonlinear sigma model is the nonlinear chiral Lagrangian based on G=H = SU(Nf )L × SU(Nf )R =SU(Nf )V for the QCD with massless Nf >avors, where Nf2 − 1 massless Nambu–Goldstone (NG) bosons are identi0ed with the pseudoscalar mesons including the meson in such an idealized limit of massless >avors. The underlying QCD dynamics generate the kinetic term of the vector mesons, which can be ignored for the energy region much lower than the vector meson mass. Then the HLS model is reduced to the nonlinear chiral Lagrangian in the low energy limit in accord with the low energy theorem of the chiral symmetry. The corresponding HLS model has the symmetry [SU(Nf )L × SU(Nf )R ]global × [SU(Nf )V ]local , with the gauge bosons of [SU(Nf )V ]local being identi0ed with the vector mesons ( meson and its >avor partners). Now, a crucial step made for the vector mesons [21–23,74] was that the vector meson mass terms were introduced in a gauge invariant manner, namely, in a way invariant under [SU(Nf )L × SU(Nf )R ]global ×[SU(Nf )V ]local and hence this mass is regarded as generated via the Higgs mechanism after gauge-0xing (unitary gauge) of HLS [SU(Nf )V ]local . 1 In writing the Gglobal × Hlocal , we had actually introduced would-be Nambu–Goldstone (NG) bosons with J PC = 0+− (denoted by , not to be confused with the scalar (so-called “sigma”) mesons having J PC = 0++ ) which are to be absorbed into the vector mesons via Higgs mechanism in the unitary gauge. Note that the usual quark >avor symmetry SU(Nf )V of QCD corresponds to H of G=H which is a residual unbroken diagonal symmetry after the spontaneous breaking of both Hlocal and Gglobal as mentioned above. The 0rst successful phenomenology was established for the and mesons in the two->avors QCD [21]: g = g

(Universality) ;

2 F2 m2 = 2g

g = 0 1

(KSRF(II)) ;

(Vector dominance) ;

(1.1) (1.2) (1.3)

There was a pre-historical work [16] discussing a concept similar to the HLS, which however did not consider a mass term of vector meson and hence is somewhat remote from the physics of vector mesons.

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for a particular choice of the parameter of the HLS Lagrangian a = 2, where g , g, m , F and g are the –– coupling, the gauge coupling of HLS, the meson mass, the decay constant of pion and the direct –– coupling, respectively. Most remarkably, we 0nd a relation independent of the Lagrangian parameters a and g [23]: g = 2g F2

(KSRF(I)) ;

(1.4)

which was conjectured to be a low energy theorem of HLS [23] and then was argued to hold at general tree-level [22]. Such a tree-level phenomenology including further developments (by the end of 1987) was reviewed in the previous Physics Reports by Bando, Kugo and one of the present authors (K.Y.) [24]. The volume included extension to the general group G and H [22], the case of generalized HLS (GHLS) Glocal , i.e., the model having the symmetry Gglobal × Glocal which can accommodate axialvector mesons (a1 meson and its >avor partners) [23,17], and the anomalous processes [74]. The success of the tree-level phenomenology is already convincing for the HLS model to be a good candidate for the e/ective 0eld theory (EFT) of the underlying QCD. It may also be useful for the QCD-like theories beyond the SM such as the technicolor [188,189,175]: the HLS model applied to the electroweak theory, sometimes called a BESS model [49,50], would be an EFT of a viable technicolor such as the walking technicolor [113,202,4,11,25] (see Refs. [200,112] for reviews) which contains the techni-rho meson. Thus the old idea of the vector meson being gauge bosons has been revived by the HLS in a precise manner: The vector meson mass is now gauge-invariant under HLS as well as invariant under the chiral symmetry of the underlying QCD. It should be mentioned that the gauge invariance of HLS does not exist in the underlying QCD and is rather generated at the composite level dynamically. This is no mystery, since the gauge symmetry is not a symmetry but simply redundancy of the description as was emphasized by Seiberg [170] in the context of duality in the SUSY QCD. Nevertheless, existence of the gauge invariance greatly simpli0es the physics as is the case in the SM. This is true even though the HLS model, based on the nonlinear sigma model, is not renormalizable in contrast to the SM. Actually, loop corrections are crucial issues for any theory of vector mesons to become an EFT and this is precisely the place where the gauge invariance comes into play. To study such loop e/ects of the HLS model as the EFT of QCD extensively is the purpose of the present Physics Reports which may be regarded as a loop version to the previous one [24]. We shall review, to the technical details, the physics of the loop calculations of HLS model developed so far within a decade in order to make the subject accessible to a wider audience. Our results may also be applicable for the QCD-like theories beyond the SM such as the technicolor and the composite W=Z models. Actually, in order that the vector meson theory be an EFT as a quantum theory including loop corrections, the gauge invariance in fact plays a vital role. It was 0rst pointed out by Georgi [85,86] that the HLS makes possible the systematic loop expansion including the vector meson loops, particularly when the vector meson mass is light. (Light vector mesons are actually realized in the Vector Manifestation which will be fully discussed in this paper.) The 0rst one-loop calculation of HLS model was made by the present authors in the Landau gauge [103] where the low energy theorem of HLS, the KSRF (I) relation, conjectured by the tree-level arguments [23,22], was con0rmed at loop level. Here we should mention [23] that being a gauge 0eld the vector meson has a de0nite o/-shell extrapolation, which is crucial to discuss the low energy theorem for the o/-shell vector

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mesons at vanishing momentum. Furthermore, a systematic loop expansion was precisely formulated in the same way as the usual chiral perturbation theory (ChPT) [190,79,81] by Tanabashi [177] who then gave an extensive analysis of the one-loop calculations in the background 0eld gauge. The low energy theorem of HLS was further proved at any loop order in arbitrary covariant gauge by Kugo and the present authors [95,96]. Also 0nite temperature one-loop calculations of the HLS was made in Landau gauge by Shibata and one of the present authors (M.H.) [102]. Here we note that there are actually many vector meson theories consistent with the chiral symmetry such as the CCWZ matter 0eld [53,48], the Massive Yang–Mills 0eld [168,169,192,77,141,128] the tensor 0eld method [79]: They are all equivalent as far as the tree-level results are concerned (see Section 3.7). However, as far as we know, the HLS model is the only theory which makes the systematic derivative expansion possible. Since these alternative models have no gauge symmetry at all, loop calculations would run into trouble particularly in the limit of vanishing mass of the vector mesons. More recently, new developments in the study of loop e/ects of the HLS were made by the present authors [104–107]: The key point was to include the quadratic divergence in the renormalization-group equation (RGE) analysis in the sense of Wilsonian RGE [195], which was vital to the chiral phase transition triggered by the HLS dynamics [104]: due to the quadratic running of F2 , the physical decay constant F (0) (pole residue of the NG bosons) can be zero, even if the bare F () de0ned at the cuto/ (just a Lagrangian parameter) is non-zero. This phenomenon supports a view [104] that HLS is an SU(Nf )-“magnetic gauge theory” dual (in the sense of Seiberg [170]) to the QCD as an SU(Nc )-“electric gauge theory”, i.e., vector mesons are “Higgsed magnetic gluons” dual to the “con0ned electric gluons” of QCD: The chiral restoration takes place independently in both theories by their respective own dynamics for a certain large number of massless >avors Nf (Nc ¡ Nf ¡ 11Nc =2), when both Nc and Nf are regarded as large [104]. Actually, it was argued in various approaches that the chiral restoration indeed takes place for the “large Nf QCD” [26,131,41,117–122,61,14,12,148,153,154,182]. The chiral restoration implies that the QCD coupling becomes not so strong as to give a chiral condensate and almost >at in the infrared region, re>ecting the existence of an infrared 0xed point (similarly to the one explicitly observed in the two-loop perturbation) and thus the large Nf QCD may be a dynamical model for the walking technicolor [113,202,4,11,25]. One might wonder why the quadratic divergences are so vital to the physics of the EFT, since as far as we do not refer to the bare parameters as in the usual renormalization where they are treated as free parameters, the quadratic divergences are simply absorbed (renormalized) into the rede0nition (rescaling) of the F2 no matter whatever value the bare F2 may take. However, the bare parameters of the EFT are actually not free parameters but should be determined by matching with the underlying theory at the cuto/ scale where the EFT breaks down. This is precisely how the modern EFT based on the Wilsonian RGE/e/ective action [195], obtained by integrating out the higher energy modes, necessarily contains quadratic divergences as physical e/ects. In such a case the quadratic divergence does exist as a physical e/ect as a matter of principle, no matter whether it is a big or small e/ect. In fact, even in the SM, which is of course a renormalizable theory and is usually analyzed without quadratic divergence for the Higgs mass squared or F (vacuum expectation value of the Higgs 0eld) renormalized into the observed value 250 GeV, the quadratic divergence is actually physical when we regard the SM as an EFT of some more fundamental theory. In the usual treatment without quadratic divergence, the bare F2 () is regarded as a free parameter and

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is freely tuned to be canceled with the quadratic divergence of order 2 to result in an observed value (250 GeV)2 , which is however an enormous 0ne-tuning if the cuto/ is physical (i.e., the SM is regarded as an EFT) and very big, say the Planck scale 1019 GeV, with the bare F2 () tuned to an accuracy of order (250 GeV)2 =(1019 GeV)2 ∼ 10−33 1. This is a famous naturalness problem, which, however, would not be a problem at all if we simply “renormalized out” the quadratic divergence in the SM. Actually, in the physics of phase transition such as in the lattice calculation, Nambu–Jona–Lasinio (NJL) model, CP N −1 , etc., as well as the SM, bare parameters are precisely the parameters relevant to the phase transition and do have a critical value due to the quadratic/power divergence, which we shall explain in details in the text. In fact, even the usual nonlinear chiral Lagrangian can give rise to the chiral symmetry restoration by the quadratic divergence of the loop [104,106]. This is actually in accord with the lattice analysis that O(4) nonlinear sigma model (equivalent to SU(2)L × SU(2)R nonlinear sigma model) give rise to the symmetry restoration for the hopping parameter (corresponding to our bare F2 ) larger than a certain critical value. The inclusion of the quadratic divergence is even more important for the phenomenological analyses when the bare HLS theory de0ned at the cuto/ scale is matched with the underlying QCD for the operator product expansion (OPE) of the current correlators (“Wilsonian matching”) [105]. Most notable feature of the Wilsonian matching is to provide the HLS theory with the otherwise unknown information of the underlying QCD such as the precise Nc -dependence which is explicitly given through the OPE. By this matching we actually determine the bare parameters of the HLS model, and hence the quadratic divergences become really physical. Most notably the bare F () is given by 2 Nc 2 ; (1.5) F () 2(1 + A ) 3 4 where A (∼ 0:5 for Nf = 3) stands for the OPE corrections to the term 1 (free quark loop). For Nc = Nf = 3 we choose 1:1 GeV ;

(1.6)

an optimal value for the descriptions of both the QCD and the HLS to be valid and the Wilsonian matching to make sense, which coincides with the naive dimensional analysis (NDA) [135], 2 ∼ 4F (0), where F (0) = 86:4 ± 9:7 MeV

(1.7)

(the “physical value” in the chiral limit mu = md = ms = 0). 3 Then we have F2 () ∼ 3(=4)2 ∼ 3(86:4 MeV)2 . Were it not for quadratic divergence, we would have predicted F2 (0) ∼ F2 () ∼ 3(86:4 MeV)2 , three times larger than the reality. It is essentially the quadratic divergence that pulls 2

The NDA does not hold for other than Nc = Nf = 3, in particular, near the chiral restoration point Nf ∼ Nfcrit with F (0) → 0 while remaining almost unchanged. For the general case other than Nc = Nf = 3 we actually 0x as (Nc =3)#s (Nc ;Nf ) = #s (3; 3 )|Nc =Nf =3 ∼ 0:7, with 3; 3 = 1:1 GeV, where #s ($) is the one-loop QCD running coupling. See Section 6.3.3. 3 This value is determined from the ratio F; phys =F (0) = 1:07 ± 0:12 given in Ref. [81], where F; phys is the physical pion decay constant, F; phys = 92:42 ± 0:26 MeV [91], and F (0) the one at the chiral limit mu = md = ms = 0. This should be distinguished from the popular “chiral limit value” 88 MeV [79] which was obtained for m2 = 0 while m2K = 0 kept to be the physical value.

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F2 down to the physical value F2 (0) ∼ 13 F2 () ∼ (86:4 MeV)2 . As to other physical quantities, the predicted values through the RGEs in the case of Nc = Nf = 3 are in remarkable agreement with the experiments [105]. It should be noted that without quadratic divergence the matching between HLS and QCD would simply break down and without vector mesons even the Wilsonian matching including the quadratic divergences would break down. When the chiral symmetry is restored in the underlying QCD with qq =0, S this Wilsonian matching 2 determines the bare parameters as a() = 1, g() = 0 and F () 2:5(Nc =3)(=4)2 = 0 (A 0:25 for qq S = 0), which we call “VM conditions” after the “vector manifestation (VM)” to be followed by these conditions. The VM conditions coincide with the Georgi’s vector limit [85,86], which, however, in contrast to the “vector realization” proposed in Refs. [85,86] with F2 (0) = 0, lead us to a novel pattern of the chiral symmetry restoration, the VM [106] with F2 (0) → 0. The VM is a Wigner realization accompanying massless degenerate (longitudinal component of) meson (and its >avor partners), generically denoted as , and the pion (and its >avor partners), generically denoted as , as the chiral partners [106]: m2 → 0 = m2 ;

F2 (0) → 0 ;

(1.8)

with m2 =F2 (0) → 0 near the critical point. The chiral restoration in the large Nf QCD can actually be identi0ed with the VM. An estimate of the critical Nf of the chiral restoration is given by a precise cancellation between the bare F2 () and the quadratic divergence (Nf =2)2 =(4)2 : 2 N f 2 Nc Nf − 0 = F2 (0) = F2 () − 2:5 ; (1.9) 2 (4)2 3 2 4 which yields Nc Nfcrit ∼ 5 (1.10) 3 in rough agreement with the recent lattice simulation [117–122], 6 ¡ Nfcrit ¡ 7 (Nc = 3) but in disagreement with that predicted by the (improved) ladder Schwinger–Dyson equation with the two-loop running coupling [14], Nfcrit ∼ 12Nc =3. Further investigation of the phase structure of the HLS model in a full parameter space leads to an amazing fact that vector dominance (VD) is no longer a sacred discipline of the hadron physics but rather an accidental phenomenon realized only for the realistic world of the Nc = Nf = 3 QCD [107]: In particular, at the VM critical point the VD is badly violated. Quite recently, it was found by Sasaki and one of the present authors (M.H.) [99] that the VM can really take place for the chiral symmetry restoration for the 0nite temperature QCD. Namely, the vector meson mass vanishes near the chiral restoration temperature in accord with the picture of Brown and Rho [42–45], which is in sharp contrast to the conventional chiral restoration aU la linear sigma model where the scalar meson mass vanishes near the critical temperature. In view of these we do believe that the HLS at loop level opened a window to a new era of the e/ective 0eld theory of QCD and QCD-like theories beyond the SM. Some technical comments are in order: In this report we con0ne ourselves to the chiral symmetric limit unless otherwise mentioned, so that pseudoscalar mesons are all precisely massless NG bosons. Throughout this report we do not include the axialvector meson (a1 meson and their >avor partners), denoted generically by A1 , since our cuto/ scale is taken as 1:1 GeV for the case

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

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Nf = 3, an optimal value where both the derivative expansion in HLS and the OPE in the underlying QCD make sense. Such a cuto/ is lower than the a1 meson mass and hence the axialvector mesons are decoupled at least for Nf = 3. If, by any chance, the axialvector mesons are to become lighter than the cuto/ near the phase transition point, our e/ective theory analysis should be modi0ed, based on the generalized HLS Lagrangian having Gglobal × Glocal symmetry [23,17]. We also omit the scalar mesons which may be lighter than the cuto/ scale [97,98,181,115,149,124] since it does not contribute to the two-point functions (current correlators) which we are studying and hence irrelevant to our analysis in this report. In this respect we note that in the HLS perturbation theory there are many counter terms (actually 35 for Nf ¿ 4) [177] compared with the usual ChPT (10 + 2 + 1 = 13) [79,81] but only few of them are relevant to the two point function (current correlators) and hence our loop calculations are reasonably tractable. It is believed according to the NDA [135] that the usual ChPT (without quadratic divergence) breaks down at the scale such that the loop correction is small: p2 2 ¡ ∼1 2 (4F (0)) (4F (0))2

(NDA) :

(1.11)

However the loop corrections generally have an additional factor Nf , i.e., Nf p2 =(4F (0))2 and hence when Nf is crucial, we cannot ignore the factor Nf . Then we should change the NDA to: [173,52] 4F (0) ∼ ; (1.12) Nf √ which yields even for Nf =3 case a somewhat smaller value ∼ 4F (0)= 3 ∼ m ¡ 1:1 GeV. This is reasonable since the appearance of pole invalidates the ChPT anyway. This is another reason why we should include in order to extend the theory to the higher scale ∼ 1:1 GeV where both the QCD (OPE) and the EFT (derivative expansion) make sense and so does the matching between them. Now, the inclusion of quadratic divergence implies that the loop corrections are given in terms of F () instead of F (0) and hence we further change the NDA to 4F () ; (1.13) ∼ Nf √ which is now consistent with the setting ∼ 1:1 GeV, since F () ∼ 3F (0) for Nf = 3 as we mentioned earlier. As to the quadratic divergence for F2 in the HLS model, the loop contributions get an extra factor 1=2 due to the additional loop, (Nf =2)p2 =(4F ())2 , and hence the loop expansion would be valid up till 4F () ∼ ; (1.14) Nf =2 which is actually the scale (or Nf when is 0xed) where the bare F2 () is completely balanced by the quadratic divergence to yield the chiral restoration F2 (0) = 0. Hence the region of the validity of the expansion is (Nf =2)2 Nf ∼ ¡1 ; 2 (4F ()) 2Nc

(1.15)

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where F2 () was estimated by Eq. (1.5) with A ∼ 0:5. This is satis0ed in the large Nc limit Nf =Nc 1, which then can be extrapolated over to the critical region Nf ∼ 2Nc . Details will be given in the text. This paper is organized as follows: In Section 2 we brie>y review the (usual) chiral perturbation theory (ChPT) [190,79,81] (without vector mesons), which gives the systematic low energy expansion of Green functions of QCD related to light pseudoscalar mesons. In Section 3 we give an up-to-date review of the model based on the HLS [21,24] at tree level. Following Ref. [24] we brie>y explain some essential ingredients of the HLS in Sections 3.2–3.5. In Section 3.6 we give a relation of the HLS to the ChPT at tree level. Section 3.7 is devoted to study the relation of the HLS to other models of vector mesons: the vector meson is introduced as the matter 0eld in the CCWZ Lagrangian [53,48] (the matter 0eld method); the massive Yang–Mills 0eld method [168,169,192,77,127,141]; and the anti-symmetric tensor 0eld method [79,70]. There we show the equivalence of these models to the HLS model. In Section 3.8, following Refs. [74,24], we brie>y review the way of incorporating vector mesons into anomalous processes, and then perform analyses on several physical processes using up-to-date experimental data. In Section 4 we review the chiral perturbation theory with HLS. First we show that, thanks to the gauge invariance of the HLS, we can perform the systematic derivative expansion with including vector mesons in addition to the pseudoscalar Nambu–Goldstone bosons in Section 4.1. The Lagrangians of O(p2 ) and O(p4 ) are given in Sections 4.2 and 4.3. In Section 4.4 we introduce the background 0eld gauge to calculate the one-loop corrections. Since the e/ect of quadratic divergences are important in this report, we explain the meaning of the quadratic divergence in our approach in Section 4.5. The explicit calculations of the two-point functions in the background 0eld gauge are performed in Section 4.6. The low energy theorem (KSRF (I)) at one-loop level is studied in Section 4.7 in the framework of the background 0eld gauge, and the renormalization group equations for the relevant parameters are given in Section 4.8. In Section 4.9 we show some examples of the relations between the parameters of the HLS and the O(p4 ) ChPT parameters following Ref. [177]. Finally in Section 4.10 we study the phase structure of the HLS following Ref. [107]. Section 5 is devoted to review the “Wilsonian matching” proposed in Ref. [105]. First, we introduce the “Wilsonian matching conditions” in Section 5.1. Then, we determine the bare parameters of the HLS using those conditions in Section 5.2 and make several physical predictions in Section 5.3. In Section 5.4 we consider QCD with Nf = 2 to show how the Nf -dependences of the physical quantities appear. Finally, in Section 5.5, we study the spectral function sum rules related to the vector and axialvector current correlators. In Section 6 we review “vector manifestation” (VM) of the chiral symmetry proposed in Ref. [106]. We 0rst explain the VM and show that it is needed when we match the HLS with QCD at the chiral restoration point in Section 6.1. Detailed characterization is also given there. Then, in Section 6.2 we review the chiral restoration in the large Nf QCD and discuss in Section 6.3 that VM is in fact be realized in the chiral restoration of the large Nf QCD. Seiberg-type duality is discussed in Section 6.4. In Section 7 we give a brief review of the proof of the low energy theorem in Eq. (1.4) at any loop order, following Refs. [95,96]. We also show that the proof is intact even when including the quadratic divergences.

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11

In Section 8 we discuss the application of the chiral perturbation with HLS to the hot and/or dense matter calculations. Following Ref. [102] we 0rst review the calculation of the hadronic thermal corrections from - and -loops in Section 8.1. In Section 8.2 following Ref. [99] we review the application of the present approach to the hot matter calculation, and in Section 8.3 we brie>y review the application to the dense matter calculation following Ref. [93]. Finally, in Section 9 we give summary and discussions. We summarize convenient formulae and Feynman rules used in this paper in Appendices A–C. A complete list of the divergent corrections to the O(p4 ) terms is shown in Appendix D. 2. A brief review of the chiral perturbation theory In this section we brie>y review the chiral perturbation theory (ChPT) [190,79,81], which gives the systematic low-energy expansion of Green functions of QCD related to light pseudoscalar mesons. The Lagrangian is constructed via non-linear realization of the chiral symmetry based on the manifold SU(Nf )L × SU(Nf )R =SU(Nf )V , with Nf being the number of light >avors. Here we generically use for the pseudoscalar NG bosons (pions and their >avor partners) even for Nf = 2. For physical pions, on the other hand, we write their charges explicitly as ± and 0 . In Section 2.1 we give a conceptual relation between the generating functional of QCD and that of the ChPT following Refs. [79,81]. Then, after introducing the derivative expansion in Section 2.2, we review how to perform the order counting systematically in the ChPT in Section 2.3. The Lagrangian of the ChPT up until O(p4 ) is given in Section 2.4. We review the renormalization and the values of the coeVcients of the O(p4 ) terms in Sections 2.5 and 2.6. The particle assignment in the realistic case of Nf = 3 is shown in Section 2.7. Finally, we review the applications of the ChPT to physical quantities such as the vector form factors of the pseudoscalar mesons (Section 2.8) and → e amplitude (Section 2.9). 2.1. Generating functional of QCD Let us start with the QCD Lagrangian with external source 0elds: LQCD = L0QCD + qSL $ L$ qL + qSR $ R$ qR + qSL [S + iP]qR + qSR [S − iP]qL ;

(2.1)

where L$ and R$ are external gauge 0elds corresponding to SU(Nf )L and SU(Nf )R , and S and P are external scalar and pseudoscalar source 0elds. L0QCD is the ordinary QCD Lagrangian with Nf massless quarks: L0QCD = qiD S , q − 12 tr[G$ G $ ] ;

(2.2)

where D$ q = (9$ − igs G$ )q ; G$ = 9$ G$ − 9 G$ − igs [G$ ; G ] ; with G$ and gs being the gluon 0eld matrix and the QCD gauge coupling constant.

(2.3)

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Transformation properties of the external gauge 0elds L and R are given by L$ → gL L$ gL† − i9$ gL · gL† ; R$ → gR R$ gR† − i9$ gR · gR† ;

(2.4)

where gL and gR are the elements of the left- and right-chiral transformations: gL; R ∈ SU(Nf )L; R . Scalar and pseudoscalar external source 0elds S and P transform as (S + iP) → gL (S + iP)gR† :

(2.5)

If there is an explicit chiral symmetry breaking due to the current quark mass, it is introduced as the vacuum expectation value (VEV) of the external scalar source 0eld:   m1   ..  : S = M =  (2.6) .   m Nf In the realistic case Nf = 3 this reads   mu    : m M= d   ms

(2.7)

Green functions associated with vector and axialvector currents, and scalar and pseudoscalar densities are generated by the functional of the above source 0elds L$ , R$ , S and P:

4 (2.8) S exp i d x LQCD : exp(iW [L$ ; R$ ; S; P]) = [dq][d q][dG] The basic concept of the ChPT is that the most general Lagrangian of NG bosons and external sources, which is consistent with the chiral symmetry, can reproduce this generating functional in the low energy region:

4 (2.9) exp(iW [L$ ; R$ ; S; P]) = [dU ] exp i d x Le/ [U; L$ ; R$ ; S; P] ; where Nf × Nf special-unitary matrix U includes the Nf2 − 1 NG-boson 0elds. In this report, for de0niteness, we use U = e2i=F ;

= a Ta ;

(2.10)

where F is the decay constant of the NG bosons . Transformation property of this U under the chiral symmetry is given by U → gL U gR† :

(2.11)

It should be noticed that the above e/ective Lagrangian generally includes in0nite number of terms with unknown coeVcients. Then, strictly speaking, we cannot say that the above generating functional agrees with that of QCD before those coeVcients are determined. Since the above generating functional is the most general one consistent with the chiral symmetry, it includes that of

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

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QCD. As one can see easily, the above generating functional has no practical use if there is no way to control the in0nite number of terms. This can be done in the low energy region based on the derivative expansion. 2.2. Derivative expansion We are now interested in the phenomenology of pseudoscalar mesons in the energy region around the mass of , p ∼ m . On the other hand, the chiral symmetry breaking scale , is estimated as [135] , ∼ 4F ∼ 1:1 GeV ;

(2.12)

where we used F = 86:4 MeV estimated in the chiral limit [79,81]. Since , is much larger than mass scale, m , , we can expand the generating functional in Eq. (2.9) in terms of m p or : (2.13) , , As is well known as Gell-Mann–Oakes–Renner relation [84], existence of the approximate chiral symmetry implies m2 ∼ M, :

(2.14)

So one can expand the e/ective Lagrangian in terms of the derivative and quark masses by assigning M ∼ O(p2 ) ; 9 ∼ O(p) : 2.3. Order counting One can show that the low energy expansion discussed in the previous subsection corresponds to the loop expansion based on the e/ective Lagrangian. Following Ref. [190], we here demonstrate this correspondence by using the scattering matrix elements of . Let us consider the matrix element with Ne external lines. The dimension of the matrix element is given by D1 ≡ dim(M ) = 4 − Ne :

(2.15)

The form of an interaction with d derivatives, k 0elds and j quark mass matrices is symbolically expressed as gd; j; k (m2 )j (9)d ()k ;

(2.16)

dim(gd; j; k ) = 4 − d − 2j − k:

(2.17)

where Let NS d; j; k denote the number of the above interaction included in a diagram for M . Then the total dimension carried by coupling constants is given by (2.18) NS d; j; k (4 − d − 2j − k) : D2 = d

j

k

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One can easily show NS d; j; k k = 2Ni + Ne ;

(2.19)

k

where Ni is the total number of internal lines. By writing NS d; j; k ; Nd; j ≡

(2.20)

k

D2 becomes Nd; j (4 − d − 2j) − 2Ni − Ne : D2 = d

(2.21)

j

By noting that the number of loops, NL , is related to Ni and Nd; j by Nd; j + 1 ; N L = Ni − d

(2.22)

j

D2 becomes D2 = 2 − 2NL + Ne +

Nd; j (2 − d − 2j) :

(2.23)

j

d

The matrix element can be generally expressed as M = E D mD 3 f(E=$; M =$) ;

(2.24)

where $ is a common renormalization scale and E is a common energy scale. The value of D3 is determined by counting the number of vertices with m : Nd; j (2j) : (2.25) D3 = d; j

D is given by subtracting the dimensions carried by the coupling constants and m from the total dimension of the matrix element M : Nd; j (d − 2) + 2NL : (2.26) D = D 1 − D2 − D3 = 2 + d; j

As we explained in the previous subsection, the derivative expansion is performed in the low energy region around the mass scale: The common energy scale is on the order of the mass, E ∼ m , and both E and m are much smaller than the chiral symmetry breaking scale , , i.e., E, m , . S is determined Then the order of the matrix element M in the derivative expansion, denoted by D, by counting the dimension of E and m appearing in M : Nd; j (d + 2j − 2) + 2NL : (2.27) DS = D + D3 = 2 + d; j

Note that N2; 0 and N0; 1 can be any number: these do not contribute to DS at all.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

15

We can classify the diagrams contributing to the matrix element M according to the value of the S Let us list examples for DS = 2 and 4. above D. 1. DS = 2 This is the lowest order. In this case, NL = 0: There are no loop contributions. The leading order diagrams are tree diagrams in which the vertices are described by the two types of terms: (d; j) = (2; 0) or (d; j) = (0; 1). Note that (d; j) = (2; 0) term includes kinetic term, and (d; j) = (0; 1) term includes mass term. 2. DS = 4 (a) NL = 1 and Nd; j = 0 [(d; j) = (2; 0); (0; 1)] One loop diagrams in which all the vertices are of leading order. (b) NL = 0 (i) N4; 0 = 1, Nd; j = 0 [(d; j) = (4; 0); (2; 0); (0; 1)] (ii) N2; 1 = 1, other Nd; j = 0 [(d; j) = (2; 1); (2; 0); (0; 1)] (iii) N0; 2 = 1, other Nd; j = 0 [(d; j) = (0; 2); (2; 0); (0; 1)] Tree diagrams in which only one next order vertex is included. The next order vertices are described by (d; j) = (4; 0), (2; 1) and (0; 2). It should be noticed that we included only logarithmic divergences in the above arguments. When we include quadratic divergences using, e.g., a method in Section 4.5, loop integrals generate the terms proportional to the cuto/ which are renormalized by the dimensionful coupling constants. 2.4. Lagrangian One can construct the most general form of the Lagrangian order by order in the derivative expansion consistently with the chiral symmetry. Below we summarize the building blocks together with the orders in the derivative expansion and the transformation properties under the chiral symmetry: U;

O(1);

U → gL UgR† ;

,;

O(p2 );

, → gL ,gR† ; , ≡ 2B(S + iP) ;

L$ ;

O(p);

L$ → gL L$ gL† − i9$ gL · gL† ;

R$ ;

O(p);

R$ → gR R$ gR† − i9$ gR · gR† ;

(2.28)

where B is a quantity of order , . Here orders of L$ and R$ are determined by requiring that all terms of the covariant derivative of U have the same chiral order: ∇$ U = 9$ U − iL$ U + iU R$ :

(2.29)

To construct the e/ective Lagrangian we need to use the fact that QCD does not break the parity as well as the charge conjugation, and require that the e/ective Lagrangian is invariant under the

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transformations under the parity (P) and the charge conjugation (C ): U↔ U† ; P

,↔ ,† ; P

L $ ↔ R$ ; P

U→ UT ; C

,↔ ,T ; C

L$ ↔ − (R$ )T ;

(2.30)

C

where the superscript T implies the transposition of the matrix. The leading order Lagrangian is constructed from the terms of O(p2 ) (DS = 2 in the previous subsection) which have the structures of (d; j) = (2; 0) or (0; 1) [79,80]: LChPT (2) =

F2 F2 tr[∇$ U † ∇$ U ] + tr[,U † + ,† U ] : 4 4

(2.31)

This leading order Lagrangian leads to the equation of motion for U up to O(p4 ): ∇$ ∇$ U † · U − U † ∇$ ∇$ U + U † , − , † U −

1 tr[U † , − ,† U ] = O(p4 ) : Nf

(2.32)

The next order is counted as O(p4 ) (DS = 4 in the previous subsection), the terms in which are described by (d; j) = (4; 0), (2; 1) or (0; 2). To write down possible terms we should note the following identities: U † ∇$ U + ∇$ U † · U = 0 ; ∇ $ U † · ∇ U + ∇ U † · ∇$ U + U † ∇ $ ∇ U + ∇ $ ∇ U † · U = 0 :

(2.33)

Now, let us list all the possible terms below: Generally, there are four terms for (d; j) = (4; 0): P0 ≡ tr[∇$ U ∇ U † ∇$ U ∇ U † ] ; P1 ≡ (tr[∇$ U † ∇$ U ])2 ; P2 ≡ tr[∇$ U † ∇ U ] tr[∇$ U † ∇ U ] ; P3 ≡ tr[∇$ U † ∇$ U ∇ U † ∇ U ] :

(2.34)

In the case of Nf = 3 we can easily show that the relation P0 = −2P3 + 12 P1 + P2

(2.35)

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17

is satis0ed. Then only three terms are independent. On the other hand, in the case of Nf = 2 the relations P0 = P2 − 12 P1 ;

P3 = 12 P1 ;

(2.36)

are satis0ed, and only two terms are independent. There are two terms for (d; j) = (2; 1): P4 ≡ tr[∇$ U † ∇$ U ] tr[,† U + ,U † ] ; P5 ≡ tr[∇$ U † ∇$ U (,† U + U † ,)] :

(2.37)

In the case of Nf = 2, we can show P5 = 12 P4 :

(2.38)

There are three terms for (d; j) = (0; 2): P6 ≡ (tr[,† U + ,U † ])2 ; P7 ≡ (tr[,† U − ,U † ])2 ; P8 ≡ tr[,† U,† U + ,U † ,U † ] :

(2.39)

In the present case there are other terms which include the 0eld strength of the external gauge 0elds L$ and R$ : P9 ≡ −i tr[L$ ∇$ U ∇ U † + R$ ∇$ U † ∇ U ] ; P10 ≡ tr[U † L$ U R$ ] :

(2.40)

In addition there are terms which include the external sources only: Q1 ≡ tr[L$ L$ + R$ R$ ] ; Q2 ≡ tr[,† ,] : One might think that there are other terms such as P˜ 1 ≡ tr[∇$ ∇$ U † · ∇ ∇ U ] :

(2.41)

However, when we want to obtain Green functions up until O(p4 ), this term is absorbed into the terms listed above by the equation of motion in Eq. (2.32) and the identity in Eq. (2.33): 1 1 1 P7 − P8 + Q2 + O(p6 ) : P˜ 1 = P3 + 4Nf 4 2

(2.42)

Namely, di/erence between the Lagrangians with and without P˜ 1 term is counted as O(p6 ) which is higher order.

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By combining the above terms the O(p4 ) Lagrangian for Nf = 3 is given by LChPT (4)

=

10

Li Pi +

i=1

2

Hi Q i

i=1

= L1 (tr[∇$ U † ∇$ U ])2 + L2 tr[∇$ U † ∇ U ] tr [∇$ U † ∇ U ] + L3 tr[∇$ U † ∇$ U ∇ U † ∇ U ] + L4 tr[∇$ U † ∇$ U ] tr[,† U + ,U † ] + L5 tr[∇$ U † ∇$ U (,† U + U † ,)] + L6 (tr[,† U + ,U † ])2 + L7 (tr[,† U − ,U † ])2 + L8 tr[,† U,† U + ,U † ,U † ] − i L9 tr[L$ ∇$ U ∇ U † + R$ ∇$ U † ∇ U ] + L10 tr[U † L$ U R$ ] + H1 tr[L$ L$ + R$ R$ ] + H2 tr[,† ,] ;

(2.43)

where Li and Hi are dimensionless parameters. Li is important for studying low energy phenomenology of the pseudoscalar mesons. For Nf = 2 case we have

LChPT (4) =

Li Pi +

i=1;2;4;6;7;8;9;10

2

Hi Qi :

(2.44)

i=1

For Nf ¿ 4 we need all the terms: LChPT (4)

=

10

Li P i +

2

i=0

Hi Qi :

(2.45)

i=1

2.5. Renormalization The parameters Li and Hi are renormalized at one-loop level. Note that all the vertices in one-loop diagrams are from O(p2 ) terms. We use the dimensional regularization, and perform the renormalizations of the parameters by Li = Lri ($) + 5i 6($);

Hi = Hir ($) + 7i 6($) ;

(2.46)

where $ is the renormalization point, and 5i and 7i are certain numbers given later. 6($) is the divergent part given by

1 1 2 − ln $ + 1 ; 6($) = − (2.47) 2(4)2 jS where 2 1 − E + ln 4 : = S j 4−n

(2.48)

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The constants 5i and 7i for Nf = 3 are given by [79,81] 51 =

3 ; 32

56 =

11 ; 144

71 = − 18 ;

52 =

3 ; 16

57 = 0; 72 =

5 24

54 = 18 ;

53 = 0; 58 =

5 ; 48

55 =

59 = 14 ;

:

3 8

;

510 = − 14 ; (2.49)

Those for Nf = 2 are given by 51 =

1 ; 12

57 = 0; 1 ; 71 = − 12

52 = 16 ;

54 = 14 ; 59 = 16 ;

58 = 0;

56 =

3 32

;

510 = − 16 ;

72 = 0 :

(2.50)

2.6. Values of low energy constants In this subsection we estimate the order of the low energy constants. By using the renormalization done just before, there is a relation between a low energy constant at a scale $ and the same constant at the di/erent scale $ : Lri ($ ) − Lri ($) =

5i $ 2 ln : 2(4)2 $2

(2.51)

If there is no accidental 0ne-tuning of parameters, we would expect the low energy constants to be at least as large as the coeVcient induced by a rescaling of order 1 in the renormalization point $. Then, Lri ($) ∼ O(10−3 )–O(10−2 ) :

(2.52)

The above estimation can be compared with the values of the low energy constants derived by 0tting to several experimental data. We show in Table 1 the values for the Nf = 3 case at $ = m9 [80] and $ = m [70]. This shows that the above estimation in Eq. (2.52) reasonably agrees with the phenomenological values of the low energy constants. 2.7. Particle assignment To perform phenomenological analyses we need a particle assignment. In a realistic case Nf = 3 there are eight NG bosons which are identi0ed with ± , 0 , K ± , K 0 , KS 0 and 9. (Strictly speaking, the octet component 98 of 9 is identi0ed with the NG boson.) These eight pseudoscalar mesons are embedded into 3 × 3 matrix as   1 0 √ + √1 98 + K+ 6 2  1    − 0 1 0 1 √ √ − + 9 K √ (2.53) =  :  6 8 2   2 K− − √26 98 KS 0

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Table 1 Values of the low energy constants for Nf = 3

Lr1 ($) Lr2 ($) Lr3 ($) Lr4 ($) Lr5 ($) Lr6 ($) Lr7 ($) Lr8 ($) Lr9 ($) Lr10 ($)

Lri ($ = m9 ) [80]

Lri ($ = m ) [70]

Source

(0:9 ± 0:3) × 10−3 (1:7 ± 0:7) × 10−3 (−4:4 ± 2:5) × 10−3 (0 ± 0:5) × 10−3 (2:2 ± 0:5) × 10−3 (0 ± 0:3) × 10−3 (−0:4 ± 0:15) × 10−3 (1:1 ± 0:3) × 10−3 (7:4 ± 0:7) × 10−3 (−6:0 ± 0:7) × 10−3

(0:7 ± 0:3) × 10−3 (1:3 ± 0:7) × 10−3 (−4:4 ± 2:5) × 10−3 (−0:3 ± 0:5) × 10−3 (1:4 ± 0:5) × 10−3 (−0:2 ± 0:3) × 10−3 (−0:4 ± 0:15) × 10−3 (0:9 ± 0:3) × 10−3 (6:9 ± 0:7) × 10−3 (−5:2 ± 0:7) × 10−3

D-waves, Zweig rule D-waves D-waves, Zweig rule Zweig rule FK : F Zweig rule Gell-Man–Okubo, L5 , L8 K 0 − K + , R, L5

r 2 e:m: → e

Values at $ = m9 is taken from Ref. [80] and those at $ = m is taken from Ref. [70].

The external gauge 0elds L$ and R$ include W$ , Z$ and A$ (photon) as g2 g2 L$ = eQA$ + (Tz − sin2 :W )Z$ + √ (W$+ T+ + W$− T− ) ; cos :W 2 R$ = eQA$ −

g2 sin2 :W Z$ ; cos :W

(2.54)

where e, g2 and :W are the electromagnetic coupling constant, the gauge coupling constant of SU(2)L and the weak mixing angle, respectively. The electric charge matrix Q is given by   2 0 0  1 0 −1 0  Q=  (2.55)   : 3 0 0 −1 Tz and T+ = (T− )†  1 1 0 Tz =  2 0

are given by  0 0  −1 0  ; 0 −1



0

 T+ =  0 0

Vud 0 0

Vus



 0   ; 0

(2.56)

where Vij are elements of Kobayashi–Maskawa matrix. 2.8. Example 1: vector form factors and L9 In this subsection, as an example, we illustrate the determination of the value of the low energy constant L9 through the analysis on the vector form factors (the electromagnetic form factors of the pion and kaon and the Kl3 form factor). We note that in the analysis of this and succeeding subsections we neglect e/ects of the isospin breaking.

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In the low energy region the electromagnetic form factor of the charged particle is given by ±

FV= (q2 ) = 1 +

1 2 =± 2 r V q + · · · ; 6

(2.57)

±

where r 2 =V is the charge radius of the particle =± and q2 is the square of the photon momentum. The electromagnetic form factor for the neutral particle is given by 0

FV= (q2 ) =

1 2 =0 2 r V q + · · · : 6

(2.58)

Similarly, one of the Kl3 form factors is given by

1 2 K 2 K 2 K f+ (q ) = f+ (0) 1 + r q + · · · ; 6

(2.59)

where r 2 K is related to the linear energy dependence 6+ by r 2 K =

66+ : m2±

(2.60)

±

±

0

In the ChPT r 2 V , r 2 KV , r 2 KV and r 2 K are calculated as [81]

m2 12Lr9 ($) 1 m2K 2 ± 2 ln 2 + ln 2 + 3 ; − r V = F2 322 F2 $ $ 0

r 2 KV = − ±

(2.61)

1 mK ln ; 2 2 16 F m 0

±

r 2 KV = r 2 V + r 2 KV ; 2 K

r

=

± r 2 V

(2.62)

2 m29 m29 5 m2K m 1 3h1 + ln 2 + 3 ln 2 − 6 ; + 3h1 − 642 F2 2 m m2K m2K mK

where 1 x3 − 3x2 − 3x + 1 1 ln x + h1 (x) ≡ 3 2 (x − 1) 2

x+1 x−1

2

−

(2.63)

1 : 3

(2.64) ±

In Ref. [80] the value of Lr9 (m9 ) is determined by using the experimental data of r 2 V given in [60]. There are several other experimental data after Ref. [80] as listed in Table 2, and they are not fully consistent. Therefore, following Ref. [81] we determine the value of Lr9 from the linear energy 0 dependence 6+ of the Ke3 form factor. By using the experimental value of 6+ given in PDG [91] 6+ = 0:0282 ± 0:0027 ;

(2.65)

the value of Lr9 (m ) is estimated as Lr9 (m ) = (6:5 ± 0:6) × 10−3 :

(2.66)

22

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

Table 2 Predictions for the charge radii of ± ; K ± K 0 in the ChPT with the existing experimental data ±

r 2 V ChPT Dally(77) [58] Molzon(78) [150] Dally(80) [59] Dally(82) [60] Amendolia(84) [7] Barkov(85) [29] Amendolia(86) [9] Amendolia(86) [8] Erkal(87) [72]

fm2

r 2 KV

0:400 ± 0:034 0:31 ± 0:04

±

0

fm2

r 2 KV fm2

0:39 ± 0:03

−0:04 ± 0:03 −0:054 ± 0:026

0:28 ± 0:05

0:439 ± 0:030 0:432 ± 0:016 0:422 ± 0:013 0:439 ± 0:008

0:34 ± 0:05 0:29 ± 0:04

0:455 ± 0:005

Using this value, we obtain the following predictions for the charge radii: ±

r 2 V = 0:400 ± 0:034 fm2 ; ±

r 2 KV = 0:39 ± 0:03 fm2 ; 0

r 2 KV = −0:04 ± 0:03 fm2 ;

(2.67) ±

±

±

0

where the error bars are estimated by [81] r 2 V =(j=2)r 2 K , r 2 KV =(j=3)r 2 V and r 2 KV = ± (j=3)r 2 V with j=±0:2. It should be noticed that the resultant charge radius of K 0 does not include any low energy constants. We show in Table 2 the comparison of the above predictions with several experimental data for the charge radii. 2.9. Example 2: → e and L10 In this subsection, we study the → e decay, and then estimate the value of the low energy constant L10 . The hadronic part is evaluated by one-pion matrix element of the vector current J$a (x) b and the axialvector current J5 (y) (a; b = 1; 2; 3) as [79]

b i d 4 x d 4 y eik ·x eip·y 0|T J$a (x)J5 (y)|c (q) · A∗$ (k)

q$ p 4(Lr9 ($) + Lr10 ($)) + (q · k g$ − q$ k ) ; = − jabc F j (k) g$ + q·k F2 ∗$

(2.68)

where A∗$ (k) is the polarization vector of the photon, A∗ (k) · k = 0. It should be noticed that the sum Lr9 ($) + Lr10 ($) is independent of the renormalization scale although each of Lr9 ($) and Lr10 ($) does depend on it. The coeVcient of the third term is related to the axialvector form factor of → ‘

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

[46,129] as F 4(Lr9 ($) + Lr10 ($)) √ A = : F 2m± By using the experimental value given by PDG [91] FA |exp = 0:0116 ± 0:0016 ;

23

(2.69)

(2.70)

the sum Lr9 ($) + Lr10 ($) is estimated as 4 Lr9 ($) + Lr10 ($) = (1:4 ± 0:2) × 10−3 :

(2.71)

By using the value of Lr9 (m ) in Eq. (2.66), Lr10 (m ) is estimated as Lr10 (m ) = (−5:1 ± 0:7) × 10−3 :

(2.72)

3. Hidden local symmetry In this section we give an up-to-date review of the model based on the hidden local symmetry (HLS) [21,24], in which the vector mesons are introduced as the gauge bosons of the HLS. Here we generically use for the pseudoscalar NG bosons (pions and their >avor partners) and for the HLS gauge bosons ( mesons and their >avor partners). We 0rst discuss the necessity for introducing the vector mesons in the e/ective 0eld theory showing a schematic view of the P-wave scattering amplitude in Section 3.1. Then, following Ref. [24] we brie>y review the model possessing the Gglobal ×Hlocal symmetry, where G =SU(Nf )L × SU(Nf )R is the global chiral symmetry and H =SU(Nf )V is the HLS, in Section 3.2. The Lagrangian of the HLS with lowest derivative terms is shown in Section 3.3 with including the external gauge 0elds. After making the particle assignment in Section 3.4, we perform the physical analysis in Section 3.5. There the parameters of the HLS are determined and several physical predictions such as the 0 → e+ e− decay width and the charge radius of pion are made. By integrating out the vector meson 0eld in the low-energy region, the HLS Lagrangian generates the chiral Lagrangian for the pseudoscalar mesons. The resultant Lagrangian is a particular form of the most general chiral perturbation theory (ChPT) Lagrangian, in which the low energy parameters Li are speci0ed. In Section 3.6 we brie>y review how to integrate out the vector mesons. Then we give predicted values of the low energy constant of the ChPT. There are models to describe the vector mesons other than the HLS. In Section 3.7 we review three models: The vector meson is introduced as the matter 0eld in the CCWZ Lagrangian [53,48] (the matter 0eld method); the massive Yang–Mills 0eld method [168,169,192,77,127,141]; and the anti-symmetric tensor 0eld method [79,70]. There we show the equivalence of these models to the HLS model. In QCD with Nf = 3 there exists a non-Abelian anomaly which breaks the chiral symmetry explicitly. In the e/ective chiral Lagrangian this anomaly is appropriately reproduced by introducing the Wess–Zumino action [193,196]. This can be generalized so as to incorporate vector mesons as the gauge bosons of the HLS [74]. We note that the low energy theorems for anomalous processes 4

For 2-loop estimation see Ref. [35].

24

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233 |T|

exp.

1-loop

tree

tree 1-loop

ρ(770)

E

Fig. 1. Schematic view of P-wave scattering amplitude.

such as 0 → 2 and → 3 are ful0lled automatically in the HLS model. In Section 3.8, following Refs. [74,24], we brie>y review the way of incorporating vector mesons, and then perform analyses on several physical processes. 3.1. Necessity for vector mesons Let us show a schematic view of the P-wave scattering amplitude in Fig. 1 [68]. As is well known, the ChPT reviewed in Section 2 explains the experimental data in the low energy region around threshold. Tree prediction of the ChPT explains the experiment in the threshold region. If we include one-loop corrections, the applicable energy region is enlarged. In the higher energy region we know the existence of meson, and the ChPT may not be applicable. So the ChPT is not so useful to explain all the data below the chiral symmetry breaking scale estimated in Eq. (2.12): , ∼ 1:1 GeV. One simple way is to include meson in the energy region. A consistent way to include the vector mesons is the HLS. Further, we can perform the similar systematic low energy expansion in the HLS as we will explain in Section 4. 3.2. Gglobal × Hlocal model Let us 0rst describe the model based on the Gglobal × Hlocal symmetry, where G = SU(Nf )L × SU(Nf )R is the global chiral symmetry and H = SU(Nf )V is the HLS. The entire symmetry Gglobal × Hlocal is spontaneously broken down to a diagonal sum H which is nothing but the H of G=H of the non-linear sigma model. This H is then the >avor symmetry. It is well known that this model is gauge equivalent to the non-linear sigma model corresponding to the coset space G=H [54–57,83,88]. The basic quantities of Gglobal × Hlocal linear model are SU(Nf )-matrix valued variables CL and CR which are introduced by dividing U in the ChPT as U = C†L CR :

(3.1)

There is an ambiguity in this division. It can be identi0ed with the local gauge transformation which is nothing but the HLS, Hlocal . These two variables transform under the full symmetry as † CL; R (x) → CL; R (x) = h(x) · CL; R (x) · gL; R ;

(3.2)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

25

where h(x) ∈ Hlocal ;

gL; R ∈ Gglobal :

(3.3)

These variables are parameterized as CL; R = ei=F e∓i=F ;

[ = a Ta ; = a Ta ] ;

(3.4)

where denote the Nambu–Goldstone (NG) bosons associated with the spontaneous breaking of G chiral symmetry and denote the NG bosons absorbed into the gauge bosons. F and F are relevant decay constants, and the parameter a is de0ned as a≡

F2 : F2

(3.5)

From the above CL and CR we can construct two Maurer–Cartan 1-forms: #⊥$ = (9$ CR · C†R 9$ CL · C†L )=(2i) ;

(3.6)

# $ = (9$ CR · C†R + 9$ CL · C†L )=(2i) ;

(3.7)

which transform as #⊥$ → h(x) · #⊥$ · h† (x) ;

(3.8)

# $ → h(x) · #⊥$ · h† (x) − i9$ h(x) · h† (x) :

(3.9)

The covariant derivatives of CL and CR are read from the transformation properties in Eq. (3.2) as D$ CL; R = 9$ CL; R − iV$ CL; R ;

(3.10)

V$ = V$a Ta

(3.11)

where are the gauge 0elds corresponding to Hlocal . These transform as V$ → h(x) · V$ · h† (x) − i9$ h(x) · h† (x) :

(3.12)

Then the covariantized 1-forms are given by 1 #ˆ⊥$ = (D$ CR · C†R − D$ CL · C†L ) ; (3.13) 2i 1 (3.14) #ˆ $ = (D$ CR · C†R + D$ CL · C†L ) : 2i The relations of these covariantized 1-forms to #⊥$ and # $ in Eqs. (3.6) and (3.7) are given by #ˆ⊥$ = #⊥$ ; #ˆ $ = # $ − V$ :

(3.15)

The covariantized 1-forms #ˆ⊥$ and #ˆ $ in Eqs. (3.13) and (3.14) now transform homogeneously: $ $ → h(x) · #⊥ · h† (x) : #⊥ ;

;

(3.16)

26

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

Thus we have the following two invariants: LA ≡ F2 tr[#ˆ⊥$ #ˆ$⊥ ] ;

(3.17)

aLV ≡ F2 tr[#ˆ $ #ˆ$ ] = F2 tr[(V$ − # $ )2 ] :

(3.18)

The most general Lagrangian made out of CL; R and D$ CL; R with the lowest derivatives is thus given by L = LA + aLV :

(3.19)

We here show that the system with the Lagrangian in Eq. (3.19) is equivalent to the chiral Lagrangian constructed via non-linear realization of the chiral symmetry based on the manifold SU(Nf )L × SU(Nf )R =SU(Nf )V , which is given by the 0rst term of Eq. (2.31) with dropping the external gauge 0elds. First, LV vanishes when we substitute the equation of motion for V$ : 5 V$ = # $ :

(3.20)

Further, with the relation 1 i #ˆ⊥$ = CL · 9$ U · C†R = CR · 9$ U † · C†L 2i 2 substituted LA becomes identical to the 0rst term of the chiral Lagrangian in Eq. (2.31):

(3.21)

F2 (3.22) tr[9$ U † 9$ U ] : 4 Let us show that the HLS gauge boson V$ agrees with Weinberg’s “-meson” [185] when we take the unitary gauge of the HLS. In the unitary gauge, = 0, two SU(Nf )-matrix valued variables CL and CR are related with each other by L = LA =

C†L = CR ≡ C = ei=F :

(3.23)

This unitary gauge is not preserved under the Gglobal transformation, which in general has the following form Gglobal : C → C = C · gR† = gL · C = exp[i (; gR ; gL )=F ] exp[i =F ] = exp[i =F ] exp[ − i (; gR ; gL )=F ] :

(3.24)

The unwanted factor exp[i (; gR ; gL )=F ] can be eliminated if we simultaneously perform the Hlocal gauge transformation with Hlocal : h = exp[i (; gR ; gL )=F ] ≡ h(; gR ; gL ) :

(3.25)

Then the system has a global symmetry G = SU(Nf )L × SU(Nf )R under the following combined transformation: G : C → h(; gR ; gL ) · C · gR† = gL · C · h† (; gR ; gL ) : 5

(3.26)

This relation is valid since we here do not include the kinetic term of the HLS gauge boson. When we include the kinetic term, this is valid only in the low energy region [see Eq. (3.91)].

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

27

Under this transformation the HLS gauge boson V$ in the unitary gauge transforms as G : V$ → h(; gR ; gL ) · V$ · h† (; gR ; gL ) − i9$ h(; gR ; gL ) · h† (; gR ; gL ) ;

(3.27)

which is precisely the same as Weinberg’s “-meson” [185]. 3.3. Lagrangian with lowest derivatives Let us now construct the Lagrangian of the HLS with lowest derivative terms. First, we introduce the external gauge 0elds L$ and R$ which include W boson, Z boson and photon 0elds as shown in Eq. (2.54). This is done by gauging the Gglobal symmetry. The transformation properties of L$ and R$ are given in Eq. (2.4). Then, the covariant derivatives of CL; R are now given by D$ CL = 9$ CL − iV$ CL + iCL L$ ; D$ CR = 9$ CR − iV$ CR + iCR R$ :

(3.28)

It should be noticed that in the HLS these external gauge 0elds are included without assuming the vector dominance. It is outstanding feature of the HLS model that CL; R have two independent source charges and hence two independent gauge bosons are automatically introduced in the HLS model. Both the vector meson 0elds and external gauge 0elds are simultaneously incorporated into the Lagrangian fully consistent with the chiral symmetry. By using the above covariant derivatives two Maurer–Cartan 1-forms are constructed as #ˆ⊥$ = (D$ CR · C†R − D$ CL · C†L )=(2i) ; #ˆ $ = (D$ CR · C†R + D$ CL · C†L )=(2i) : These 1-forms are expanded as 1 i 1 9$ + A $ − [V$ ; ] − [[9$ ; ]; ] + · · · ; #ˆ⊥$ = F F 6F3 1 i i 9$ − V$ + V$ − [9$ ; ] − [A$ ; ] + · · · ; #ˆ $ = F 2F2 F

(3.29)

(3.30) (3.31)

where V$ = (R$ + L$ )=2 and A$ = (R$ − L$ )=2. The covariantized 1-forms in Eqs. (3.29) transform homogeneously: #ˆ$ ; ⊥ → h(x) · #ˆ$ ; ⊥ · h† (x) :

(3.32)

Then we can construct two independent terms with lowest derivatives which are invariant under the full Gglobal × Hlocal symmetry as LA ≡ F2 tr[#ˆ⊥$ #ˆ$⊥ ] = tr[9$ 9$ ] + · · · ;

(3.33)

aLV ≡ F2 tr[#ˆ $ #ˆ$ ] = tr[(9$ − F V$ )(9$ − F V $ )] + · · · ;

(3.34)

where the expansions of the covariantized 1-forms in Eqs. (3.30) and (3.31) were substituted to obtain the second expressions. These expansions imply that LA generates the kinetic term of pseudoscalar

28

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

meson, while LV generates the kinetic term of the would-be NG boson in addition to the mass term of the vector meson. Another building block is the gauge 0eld strength of the HLS gauge boson de0ned by V$ ≡ 9$ V − 9 V$ − i[V$ ; V ] ;

(3.35)

which also transforms homogeneously: V$ → h(x) · V$ · h† (x) :

(3.36)

Then a simplest term with V$ is the kinetic term of the gauge boson: 1 Lkin (V$ ) = − 2 tr[V$ V $ ] ; 2g

(3.37)

where g is the HLS gauge coupling constant. Now the Lagrangian with lowest derivatives is given by [21,24] L = LA + aLV + Lkin (V$ ) = F2 tr[#ˆ⊥$ #ˆ$⊥ ] + F2 tr[#ˆ $ #ˆ$ ] −

1 tr[V$ V $ ] : 2g2

(3.38)

3.4. Particle assignment Phenomenological analyses are performed with setting Nf = 3 and extending the HLS to Hlocal = [U(3)V ]local . Accordingly, the chiral symmetry is extended to Gglobal = [U(3)L × U(3)R ]global . Then the pseudoscalar meson 0eld matrix becomes =

8 a=0

Ta a 

1   =√  2

0 + 2

√1

√1

9 + 6 8

− K−

+

√1

9 3 0 − √12 0 +

√1

9 + 6 8



K+ √1

9 3 0

KS 0

K0 − √26 98 +

√1

3

90

   ; 

(3.39)

where appropriate combinations of 98 and 90 become 9 and 9 . The HLS gauge boson 0eld matrix is expressed as V$ =

8

Ta V$a ;

a=0

1 T0 = √ : 6

(3.40)

Strictly speaking, we need to introduce the e/ect of the violation of Okubo–Zweig–Iizuka (OZI) rule [155,204,205,116] when we perform the systematic low-energy expansion. That e/ect is expressed by the following Lagrangian: LOZIB; (2) =

F;2 B F;2 B 1 tr[#ˆ⊥$ ] tr[#ˆ$⊥ ] + tr[#ˆ $ ] tr[#ˆ$ ] − tr[V$ ] tr[V $ ] : Nf Nf 2Nf gB2

(3.41)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

29

However, it is well known that the OZI rule works very well for vector meson nonet. Then it is natural to take 6 F; B = 0;

1 =0 : gB

(3.42)

In such a case, it is convenient to introduce the following nonet:  √1 (0 + !$ ) + $ $ 2  1  − 0 1 $ − √2 ($ + !$ ) $ = V$ =g = √  2 K$∗; − KS ∗$; 0

particle assignment for the vector meson K$∗; + ∗; 0

K$

   ; 

(3.43)

=$

where we used the ideal mixing scheme:   8 1 2 V$ =g !$ 3 3   : =  =$ V$0 =g 1 − 2 3



(3.44)

3

The embedding of W$ , Z$ and A$ (photon) in the external gauge 0elds L$ and R$ were done in Section 2.7. Here, just for convenience, we list it again: g2 g2 L$ = eQA$ + (Tz − sin2 :W )Z$ + √ (W$+ T+ + W$− T− ) ; cos :W 2 R$ = eQA$ −

g2 sin2 :W Z$ ; cos :W

(3.45)

where e, g2 and :W are the electromagnetic coupling constant, the gauge coupling constant of SU(2)L and the weak mixing angle, respectively. The electric charge matrix Q is given by   2 0 0  1  : 0 −1 0 Q=  (3.46)  3 0 0 −1 Tz and T+ = (T− )†  1 1 0 Tz =  2 0

are given by   0 0 0    −1 0   ; T+ =  0 0 0 −1

Vud 0 0

Vus



 0   ; 0

(3.47)

where Vij are elements of Kobayashi–Maskawa matrix. 6

Note that OZI violating e/ect to the pseudoscalar meson decay constant is needed for phenomenological analysis (see, e.g., Ref. [167]).

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

3.5. Physical predictions at tree level Let us study some phenomena using the Lagrangian with lowest derivatives given in Eq. (3.38). In this Lagrangian all the vector mesons are degenerate even when we apply the HLS to the case of Nf = 3. The mass splitting among the vector-meson nonet (or octet) is introduced when we include the higher derivative terms (see Section 4). So we study some phenomenology related to the meson. By taking the unitary gauge of the HLS ( = 0) and substituting the expansions of #ˆ⊥$ and #ˆ $ given in Eqs. (3.30) and (3.31) into the Lagrangian in Eq. (3.38), we obtain L = tr[(9$ − i[A$ Q; ] + · · ·)2 ] 2 i 2 (3.48) g$ − eA$ Q + [9$ ; ] + · · · + aF tr 2F2 1 ag tr[$ [9$ ; ]] = tr[9$ 9$ ] + ag2 F2 tr[$ $ ] + 2i 2 a $ A tr[Q[9$ ; ]] − 2eagF2 A$ tr[$ Q] + 2ie 1 − 2 4 − 3a tr[[9$ ; ][9$ ; ]] + · · · ; (3.49) + ae2 F2 A$ A$ tr[QQ] + 12F2 where we have gauged only a subgroup of Gglobal , Iglobal =U(1)Q ⊂ Hglobal ⊂ Gglobal =SU(3)L ×SU(3)R , with the photon 0eld A$ in Eq. (3.45), and the vector meson 0eld $ related to V$ by rescaling the kinetic term in Eq. (3.37): V$ = g$ :

(3.50)

From this we can easily read the meson mass m , the coupling constant g , the – mixing strength g and the direct coupling constant g : m2 = ag2 F2 ;

(3.51)

1 ag ; (3.52) 2 g = agF2 ; (3.53) a e : (3.54) g = 1 − 2 We should note that the acquires a mass through the Higgs mechanism associated with spontaneous breaking of the HLS Hlocal . We also note that the photon denoted by A$ in Eq. (3.49) also acquire the mass through the Higgs mechanism since the photon is introduced by gauging the subgroup Iglobal = U(1)Q ⊂ Gglobal which is spontaneously broken together with the HLS. Thus Hlocal × (gauged-)Iglobal → U(1)em . When we add the kinetic term of the photon 0eld A$ in the Lagrangian in Eq. (3.49), the photon 0eld mixes with the neutral vector meson (0 for Nf = 2). For Nf = 2 the mass matrix of the photon and 0 are given by 0$ 2 eg g aF2 (0$ ; A$ ) ; (3.55) 2 A$ eg e g =

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

31

which is diagonalized by introducing new 0elds de0ned by 1 ˜0$ ≡ (g0$ − eA$ ) ; 2 g + e2 1 A˜ $ ≡ (g0$ + eA$ ) : 2 g + e2

(3.56)

The mass eigenvalues are given by m2˜ 0 = aF2 (g2 + e2 ) ;

(3.57)

m2A˜ = 0 :

(3.58)

m2± = aF2 g2 :

(3.59)

The charged vector mesons ± of course do not mix with the photon, and the masses of them are given by For Nf = 2 the above situation implies that the [SU(2)V ]HLS × U(1)Q symmetry is spontaneously broken down to U(1)em . The massless gauge boson of the remaining U(1)em is nothing but the physical photon 0eld A˜ $ in Eq. (3.56). This situation is precisely the same as that occurring in the Glashow–Salam–Weinberg model. Comparing the mass of neutral in Eq. (3.57) with the mass of charged in Eq. (3.59), we immediately conclude that the neutral is heavier than the charged : m˜ 0 ¿ m± . Furthermore, we have the following prediction for the mass di/erence between the neutral and the charged : e2 √ m˜ 0 − m± aF ∼ 1 MeV ; (3.60) 2g where we used e2 = 4=137 0:092; F 92 MeV [see Eq. (3.66)], g 5:8 [see Eq. (3.74)] and a 2:1 [see Eq. (3.75)]. For Nf = 2 the above mass di/erence in Eq. (3.60) is consistent with the experimental value of the 0 -± mass di/erence [91]: m0 − m± |exp = 0:5 ± 0:7 MeV :

(3.61)

Future experiment is desirable for checking the prediction (3.60) of the HLS. Now we turn to a discussion of the implication of the relations among the masses and coupling constants in Eqs. (3.51)–(3.54). For a parameter choice a = 2, the above results reproduce the following outstanding phenomenological facts [21]: (1) g = g (universality of the -coupling) [165] 2 (2) m2 = 2g F2 (KSRF II) [126,163] (3) g = 0 ( meson dominance of the electromagnetic form factor of the pion) [165]. Moreover, independently of the parameter a, Eqs. (3.52) and (3.53) lead to the KSRF relation [126,163] (version I) g = 2F2 g :

(3.62) $

This parameter independent relation comes from the ratio of the two cross terms $ A and $ [9$ ; ] in aLV term [second term in Eq. (3.48)], so that it is obviously independent of a which is an overall

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

factor. Moreover, the ratio is precisely 0xed by the symmetry Gglobal × Hlocal of our Lagrangian with the subgroup Iglobal ⊂ Gglobal being gauged, and hence is a direct consequence of the HLS independently of dynamical details. Since the o/-shell extrapolation of the vector meson 0elds are well de0ned in the HLS, the KSRF (I) relation also makes sense for the o/-shell at soft momentum limit: 2 2 g (p2 = 0) = 2g (p2 = 0; q1 = 0; q2 = 0)F2 ;

(3.63)

where p is the momentum and q1 and q2 are the pion momenta. This relation is actually a low-energy theorem of the HLS [23] to be valid independently of any higher derivative terms which are irrelevant to the low-energy limit. This low-energy theorem was 0rst proved at the tree level [22], then at one-loop level [103] and any loop order [95,96]. Importance of this low-energy theorem is that although it is proved only at the low-energy limit, the KSRF (I) relation actually holds even at the physical point on the mass-shell. g and g in Eq. (3.62) are related to the → decay width and the → e+ e− decay width as 3 m2 − 4m2 |˜ p | 2 ; (3.64) 5( → ) = |g | ; |˜ p | = 6m2 4 4#2 g 2 m2 + 2m2e 2 + − 5( → e e ) = m − 4m2e : (3.65) 3 m2 m2 By using the experimental values [91] F = 92:42 ± 0:26 MeV ;

(3.66)

m = 771:1 ± 0:9 MeV ;

(3.67)

m = 139:57018 ± 0:00035 MeV ;

(3.68)

5( → )exp = 149:2 ± 0:7 MeV ;

(3.69)

5( → e+ e− )exp = 6:85 ± 0:11 keV ;

(3.70)

the values of g and g are estimated as g |exp = 6:00 ± 0:01 ;

(3.71)

g |exp = 0:119 ± 0:001 GeV2 :

(3.72)

From these experimental values we obtain g = 1:15 ± 0:01 : 2g F2 exp

(3.73)

This implies that the KSRF (I) relation in Eq. (3.62) is well satis0ed, which may be regarded as a decisive test of the HLS. 7 The above small deviation of the experimental values from the KSRF (I) 7

When we use 5( → ) as an input and predict the → e+ e− decay width from the low-energy theorem, we obtain 5( → e+ e− ) = 5:11 ± 0:23 keV.

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33

relation is on-shell corrections due to the non-zero mass. Actually, as we shall show in Section 5, the di/erence of the value in Eq. (3.73) from one is explained by the corrections from the higher derivative terms. Now, let us determine three parameters F , a and g from the experimental data. The value of F is just taken from the experimental value in Eq. (3.66). We determine the values of a and g from g |exp in Eq. (3.71) and m in Eq. (3.67) through Eqs. (3.51) and (3.52). Then the values of the parameters a and g are determined as 8 2 m2 g = 2:67 ± 0:84 ; (3.74) g= = 5:80 ± 0:91 2g F2 4 a=

2g = 2:07 ± 0:33 ; g

(3.75)

where we add 15% error for each parameter, which is expected from the deviation of the low-energy theorem in Eq. (3.73). From these values the predicted value of g is given by g = 0:103 ± 0:023 GeV2 ;

(3.76)

which is compared with the value in Eq. (3.72) obtained from 5( → e+ e− ). Before making physical predictions, let us see the electromagnetic form factor of the pion. If one sees only the direct coupling in Eq. (3.54), one might think that the electric charge of would not be normalized to be unity, and thus the gauge invariance of the photon would be violated. This is obviously not the case, since the Lagrangian (3.38) or (3.48) is manifestly gauge invariant under U(1)em by construction. This can also be seen diagrammatically as follows. The term proportional to a in g of Eq. (3.54) comes from the vertex derived from aLV term in the Lagrangian (3.38), and then it is exactly canceled with the -exchange contribution coming from the same aLV term in the low energy limit. Thus, the electric charge of is properly normalized. To visualize this, we show the diagrams contributing to the electromagnetic form factor of ± in Fig. 2. The contributions 8

If we determine g and a from g of Eq. (3.72) and m of Eq. (3.67), then we have 2 m2 g = 5:01 ± 0:79 g= = 2:00 ± 0:63 ; g 4 a=

g = 2:77 ± 0:44 : gF2

There is about 15% di/erence between the above value of g and that in Eq. (3.74), as implied by Eq. (3.73). Since Eqs. (3.51)–(3.53) lead to a=

2 4g g2 F2 = 2 2 ; 2 m m F

there is about 30% di/erence between the above value of a and that in Eq. (3.75). One might think that we could use the above values of g and a for phenomenological analysis. However, as we will show in Section 5, the deviation of the prediction of g in Eq. (3.76) from the experimental value in Eq. (3.72) is explained by including the higher derivative term (z3 term). Thus, we think that it is better to use the values in Eqs. (3.74) and (3.75) for the phenomenological analysis at tree level. Actually, the values of g and a in Eqs. (3.74) and (3.75) are consistent with those obtained by the analysis based on the Wilsonian matching as shown in Section 5. [See g(m ) in Table 8 and a(0) in Table 9.]

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q2

π

γ

+ q1

+

ρ

π

(a)

(b)

(c)

Fig. 2. Electromagnetic form factor in the HLS: (a) the direct interaction from LA term in the Lagrangian (3.38); (b) the direct interaction from aLV term; (c) the interaction mediated by exchange.

from the diagrams in Fig. 2 are summarized as 5$(a) (q2 ; q1 ) = e(q1 + q2 )$ ; a ; 5$(b) (q2 ; q1 ) = e(q1 + q2 )$ − 2 g g 5$(c) (q2 ; q1 ) = e(q1 + q2 )$ 2 ; m − p 2

(3.77)

where p$ = q2$ − q1$ . By summing these contributions with noting the relation g g = am2 =2, the electromagnetic form factor of ± is given by ±

FV (p2 ) = 1 −

m2 a a + : 2 2 m2 − p2

(3.78)

In this form we can easily see that the contributions from the diagrams (b) and (c) in Fig. 2 are exactly canceled in the low energy limit p2 = 0, and thus the electromagnetic form factor of pion is properly normalized: ±

FV (p2 = 0) = 1 :

(3.79)

Now, we make physical predictions using the values of the parameters in Eqs. (3.66), (3.74) and ± (3.75). An interesting physical quantity is the charge radius of pion r 2 V , which is de0ned through the electromagnetic form factor of ± in the low energy region as p2 2 ± (3.80) r V + · · · : 6 From the electromagnetic form factor in Eq. (3.78), which is derived from the Lagrangian with lowest derivatives in Eq. (3.38), the charge radius of ± is expressed as g g 3a ± r 2 V = 6 = 2 : (3.81) 4 m m ±

FV (p2 ) = 1 +

By using the value of a in Eq. (3.75) and the experimental value of meson mass in Eq. (3.67) this is evaluated as ±

r 2 V = 0:407 ± 0:064 fm2 :

(3.82)

Comparing this with the experimental values shown in Table 2 in Section 2, we conclude that the HLS model with lowest derivatives reproduces the experimental data of the charge radius of pion very well.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

+ (a)

+ (b)

35

ρ (c)

Fig. 3. Diagrams contributing to the scattering in the HLS: (a) contribution from the contact 4-interaction from LA term in the Lagrangian (3.38); (b) contribution from the contact 4-interaction from aLV term; (c) contribution from the -exchange. The diagram (c) implicitly includes three diagrams: s-channel, t-channel and u-channel -exchange diagrams.

Another interesting physical quantity is the axialvector form factor FA of → ‘ studied in Section 2.9. In the HLS with lowest derivatives there is no contribution to this axialvector form factor, and thus FA = 0. This, of course, does not agree with the experimental data in Eq. (2.70). However, as we shall show in Section 5, the prediction of the HLS reasonably agree with the experiment when we go to the next order, O(p4 ). Finally in this subsection, we consider the low-energy theorem on the scattering amplitude, which is a direct consequence of the chiral symmetry. If one sees the contact 4-interaction in Eq. (3.49), one might think that the HLS violated the low-energy theorem of the scattering amplitude. However, this is of course not true since the Lagrangian (3.38) is chiral-invariant and hence must respect the low-energy theorem trivially. This can be also seen diagrammatically as follows: The term proportional to a in the contact 4-interaction is derived from aLV term in the Lagrangian (3.38), which is exactly canceled by the -exchange contribution in the low-energy limit. To visualize this, we show the diagrams contributing to the scattering in Fig. 3. Contributions to the scattering amplitude A(s; t; u) are given by 9 s A(a) (s; t; u) = 2 ; (3.83) F 3as (3.84) A(b) (s; t; u) = − 2 ; 4F

u−s t−s 2 : (3.85) A(c) (s; t; u) = −g + m2 − t m2 − u 2 Noting that a=(4F2 ) = g =m2 , we obtain

2 g t(u − s) u(t − s) (b+c) + 2 ; (s; t; u) = − 2 A m m2 − t m − u

(3.86)

where we used s + t + u = 0. Thus, the sum of the contributions from (b) and (c) does not contribute in the low-energy limit and only the diagram (a) contributes, which is perfectly consistent with the low-energy theorem of the -scattering amplitude. This can be easily seen as follows: In the low-energy region we can neglect the kinetic term of , i.e., Lkin (V$ ) = 0 in Eq. (3.38), and then the 0eld V$ becomes just an auxiliary 0eld. Integrating out the auxiliary 0eld V$ leads to aLV = 0 9

The invariant amplitude for i (p1 ) + j (p2 ) → k (p3 ) + l (p4 ) is decomposed as ij kl A(s; t; u) + ik jl A(t; s; u) + il jk A(u; t; s), where s; t and u are the usual Mandelstam variables: s = (p1 + p2 )2 ; t = (p1 + p3 )2 and u = (p1 + p4 )2 .

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in Eq. (3.38). There remains only LA term which is nothing but the chiral Lagrangian with the least derivative term. Then the result precisely reproduces the low-energy theorem. 3.6. Vector meson saturation of the low energy constants (relation to the ChPT) Integrating out the vector mesons in the Lagrangian of the HLS given in Eq. (3.38) we obtain the Lagrangian for pseudoscalar mesons. The resultant Lagrangian includes O(p4 ) terms of the ChPT in addition to O(p2 ) terms. To perform this it is convenient to introduce the following quantities: #⊥$ = (D$ CR · C†R − D$ CL · C†L )=(2i) ; # $ = (D$ CR · C†R + D$ CL · C†L )=(2i) ;

(3.87)

where D$ CL and D$ CR are de0ned by D$ CL = 9$ CL + iCL L$ ; D$ CR = 9$ CR + iCR R$ :

(3.88)

The relations of these #⊥$ and # $ with #ˆ⊥$ and #ˆ $ in Eq. (3.29) are given by #ˆ⊥$ = #⊥$ ; #ˆ $ = # $ − V$ :

(3.89)

From the Lagrangian in Eq. (3.38) the equation of motion for the vector meson is given by 1 (3.90) F2 (V$ − # $ ) − 2 (9 V$ − i[V ; V$ ]) = 0 : g In the leading order of the derivative expansion the solution of Eq. (3.90) is given by 1 V$ = # $ + 2 O(p3 ) : (3.91) m Substituting this into the 0eld strength of the HLS gauge boson and performing the derivative expansion we obtain ˆ$ + i[#ˆ⊥$ ; #ˆ⊥ ] + 1 O(p4 ) V$ = V m2 i i 1 = CL U R$ U † + L$ + ∇$ U · ∇ U † − ∇ U · ∇$ U † C†L + 2 O(p4 ) 4 4 m i i 1 † † † = CR R$ + U L$ U + ∇$ U · ∇ U − ∇ U · ∇$ U C†R + 2 O(p4 ) ; (3.92) 4 4 m where we used i 1 (3.93) #ˆ⊥$ = CL · ∇$ U · C†R = CR · ∇$ U † · C†L : 2 2i By substituting Eq. (3.93) into the HLS Lagrangian, the 0rst term in the HLS Lagrangian (3.38) becomes the 0rst term in the leading order ChPT Lagrangian in Eq. (2.31): F2 tr[∇$ U † ∇$ U ] : LChPT (3.94) (2) |,=0 = 4

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37

Table 3 Values of low-energy constants derived from the HLS Lagrangian with lowest derivatives

L1 L2 L3 L4 L5 L6 L7 L8 L9 L10

Lri (m ) × 103

LVi × 103

0:7 ± 0:3 1:3 ± 0:7 −4:4 ± 2:5 −0:3 ± 0:5 1:4 ± 0:5 −0:2 ± 0:3 −0:4 ± 0:15 0:9 ± 0:3 6:9 ± 0:7 −5:2 ± 0:7

0:93 ± 0:29 1:86 ± 0:58 −5:6 ± 1:8

7:4 ± 2:3 −7:4 ± 2:3

In addition, the second term in Eq. (3.38) with Eq. (3.90) substituted becomes of O(p6 ) in the ChPT and the third term (the kinetic term of the HLS gauge boson) with Eq. (3.92) becomes of O(p4 ) in the ChPT: LV4 =

1 1 (tr[∇$ U ∇$ U † ])2 + tr[∇$ U ∇ U † ] tr[∇$ U ∇ U † ] 2 32g 16g2 −

3 1 tr[∇$ U ∇$ U † ∇ U ∇ U † ] − i 2 tr[L$ ∇$ U ∇ U † + R$ ∇$ U † ∇ U ] 2 16g 4g

−

1 1 tr[L$ U R$ U † ] − 2 [L$ L$ + R$ R$ ] ; 2 4g 8g

(3.95)

where we 0xed Nf = 3. Comparing this with the O(p4 ) terms of the ChPT Lagrangian given in Eq. (2.43), we obtain the contributions of vector mesons to the low-energy parameters of the ChPT: LV1 =

1 ; 32g2

LV2 =

1 ; 16g2

LV9 =

1 ; 4g2

LV10 = −

1 : 4g2

LV3 = −

3 ; 16g2 (3.96)

In Table 3 we show the values of LVi obtained by using the value of g determined in the previous subsection, g = 5:80 ± 0:91 [Eq. (3.74)], with the values of Lri (m ) in Ref. [70]. This shows that the low-energy constants L1 ; L2 ; L3 and L9 are almost saturated by the contributions from vector mesons at the leading order [70,71,68]. L10 will be saturated by including the next order correction (see Section 5).

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3.7. Relation to other models of vector mesons There are models to describe the vector mesons other than the HLS. In this subsection, we introduce several models of the vector mesons, and show the equivalence between those and the HLS. In Ref. [71] it was shown that the vector meson 0eld can be introduced as the matter 0eld in the CCWZ Lagrangian [53,48]. Hereafter we call this model the matter 0eld method. The equivalence of the model to the HLS was studied in Refs. [71,36]. However, the higher order terms of the HLS, which we will show in Section 4, were not considered. Here we show the equivalence including the higher order terms in the HLS after brie>y reviewing the matter 0eld method. Another popular model is the so-called “Massive Yang–Mills” 0eld method [168,169,192,77, 127,141]. Although the notion of the “Massive Yang–Mills” itself does not literally make sense due to the mass term introduced by hand, the real meaning of “Massive Yang–Mills” approach was revealed [198,24] in terms of the generalized HLS (GHLS) including the axialvector mesons [23,17]: The “Massive Yang Mills” Lagrangian is nothing but a special gauge of the GHLS with a particular parameter choice and hence equivalent to the HLS model after eliminating the axialvector mesons [166,198,89,143]. (For reviews, see Refs. [24,141].) We here brie>y review the equivalence to the “Massive Yang–Mills” model in view of GHLS. In Refs. [79,70] the vector mesons are introduced as anti-symmetric tensor 0elds. The equivalence was studied in Refs. [71,178]. Especially in Ref. [178], the equivalence was shown with including the higher order terms of the HLS. Here we brie>y review the model and equivalence mostly following Ref. [178]. In the following discussions we restrict ourselves to the chiral limit. The extensions to the case with the explicit chiral symmetry breaking by the current quark masses are automatic. As we will show below, there are di/erences in the o/-shell amplitude since the de0nitions of the o/-shell 0elds are di/erent in the models. Moreover, we can make the systematic derivative expansion in the HLS as we will show in Section 4, while we know no such systematic expansions in other models. Thus, the equivalence is valid only for the tree level on-shell amplitude. 3.7.1. Matter ?eld method Let us show the equivalence between the matter 0eld method and the HLS. We 0rst brie>y describe the nonlinear sigma model based on the manifold G=H [53,48] with restricting ourselves to the case for G = SU(Nf )L × SU(Nf )R and H = SU(Nf )V , following Ref. [24]. 10 Let C() be “representatives” of the (left) coset space G=H , taking the value of the unitary matrix representation of G, which are conveniently parametrized in terms of the NG bosons (x) as C() = ei(x)=F ;

(x) = a (x)Ta ;

(3.97)

where we omit the summation symbol over a. The transformation property of C() under the chiral symmetry is given by G : C() → C( ) = h(; gR ; gL ) · C() · gR† = gL · C() · h† (; gR ; gL ) : 10

An explanation in the present way for general G and H was given in Ref. [24].

(3.98)

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39

The fundamental objects are the following Maurer–Cartan 1-forms constructed from C() ∈ G=H : 11 1 $ [D C · C† − D$ C† · C]; 2i 1 # $ = [D$ C · C† + D$ C† · C] ; 2i $ #⊥ =

(3.99)

where D$ C and D$ C† are de0ned by D$ C† ≡ 9$ C† + iC† L$ ; D$ C ≡ 9$ C + iCR$ :

(3.100)

The transformation properties of these 1-forms are given by $ $ #⊥ → h(; gR ; gL ) · #⊥ · h† (; gR ; gL ) ;

# $ → h(; gR ; gL ) · # $ · h† (; gR ; gL ) − i9$ h(; gR ; gL ) · h† (; gR ; gL ) :

(3.101)

$ transforms homogeneously, so that we can construct G-invariant Only the perpendicular part #⊥ $ from #⊥ alone: $ #⊥ $ ] ; LCCWZ = F2 tr[#⊥

(3.102)

where the factor F2 is added so as to normalize the kinetic terms of the (x) 0elds. It should be noticed that with the relation #⊥$ =

1 † i C · ∇ $ U · C† = C · ∇ $ U † · C ; 2i 2

(3.103)

substituted LCCWZ becomes identical to the 0rst term of the chiral Lagrangian in Eq. (2.31). Following Ref. [71], we include the vector meson as the matter 0eld in the adjoint representation, = (C)a Ta : (3.104) (C) $ $ a

This transforms homogeneously under the chiral symmetry: (C) † (C) $ → h(; gR ; gL ) · $ · h (; gR ; gL ) :

(3.105)

The covariant derivative acting on the vector meson 0eld is de0ned by (C) (C) D$(C) (C) ≡ 9$ − i[# $ ; ] :

(3.106)

It is convenient to de0ne the following anti-symmetric combination of the above covariant derivative: (C) (C) (C) (C) (C) : $ ≡ D$ − D $

11

(3.107)

In Refs. [70,71] u$ and 5$ were used instead of #⊥$ and #$ . The relations between them are given by u$ = 2#⊥$ and 5$ = −i#$ .

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In addition we need the 0eld strengths of the external source 0elds L$ and R$ . These are given by ˆ$ ≡ 1 [CR$ C† + C† L$ C] ; V 2 1 Aˆ$ ≡ [CR$ C† − C† L$ C] ; 2

(3.108)

which transform homogeneously: ˆ$ → h(; gR ; gL ) · V ˆ$ · h† (; gR ; gL ) ; V Aˆ$ → h(; gR ; gL ) · Aˆ$ · h† (; gR ; gL ) :

(3.109)

ˆ$ and Aˆ$ agree with those in Eq. (4.24) when the unitary gauge Note that these expressions of V (C) ˆ ˆ of the HLS is taken. The above #⊥$ ; (C) $ ; $ ; V $ and A$ together with the covariant derivative acting these 0elds de0ned by D$(C) ≡ 9$ − i[# $ ; ] ;

(3.110)

are the building blocks of the Lagrangian of the matter 0eld method. The Lagrangian of the matter 0eld method is constructed from the building blocks given above. An example of the Lagrangian including the vector meson is given by [71] 1 (C)$ (C)$ LC = − tr[(C) ] + M2 tr[(C) ] $ $ 2 (C) $ ˆ $ − fV tr[(C) $ V ] − 4igV tr[$ #⊥ #⊥ ] :

(3.111)

In order to make the procedure more systematic, the terms including the pseudoscalar meson are added to the Lagrangian in Eq. (3.111) in Ref. [71]. The entire Lagrangian is given by (C) (3.112) LCS = LCCWZ + LC + i Pi ; i=1;2;3;9;10

where Pi is the O(p4 ) terms in the ChPT de0ned in Eqs. (2.34), (2.37), (2.39) and (2.40). By using #⊥$ these Pi (i = 1; 2; 3; 9; 10) are expressed as $ 2 ]) ; P1 = 16 tr([#⊥$ #⊥ $ P2 = 16 tr[#⊥$ #⊥ ] tr[#⊥ #⊥ ] ; $ P3 = 16 tr[#⊥$ #⊥ #⊥ #⊥ ] ; ˆ$ #$ #⊥ P9 = −8i tr[V ] ; ⊥

ˆ$ V ˆ $ ] − tr[Aˆ$ Aˆ$ ] : P10 = tr[V

(3.113)

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41

Now that we have speci0ed the Lagrangian for the matter 0eld method, we compare this with the HLS Lagrangian. This is done by rewriting the above vector meson 0eld (C) into #ˆ $ of the $ HLS as (C) $ = I#ˆ $ = I(# $ − V$ ) ;

(3.114)

where I is a parameter related to the rede0nition of the vector meson 0eld V$ in the HLS. It should be noticed that this relation is valid only when we take the unitary gauge of the HLS. The covariant derivative D$(C) is related to that in the HLS D$ as D$(C) = 9$ − i[V$ ; ] − i[#ˆ $ ; ] = D$ − i[#ˆ $ ; ] : Then

(C) $

in Eq. (3.107) is rewritten as

(C) $

= I(D$ #ˆ − D #ˆ $ − 2i[#ˆ $ ; #ˆ ]) ˆ$ − IV$ ; = −iI[#ˆ $ ; #ˆ ] + iI[#ˆ⊥$ ; #ˆ⊥ ] + IV

(3.115)

(3.116)

where D$ #ˆ ≡ 9$ #ˆ − i[V$ ; #ˆ ] ; and to obtain the second expression we used the following identity [see Eq. (4.32)]: ˆ$ − V$ : D$ #ˆ − D #ˆ $ = i[#ˆ $ ; #ˆ ] + i[#ˆ⊥$ ; #ˆ⊥ ] + V

(3.117) (3.118)

In addition, as shown in Eq. (3.89) #⊥$ agrees with #ˆ⊥$ in the unitary gauge of the HLS: #⊥$ = #ˆ⊥$ :

(3.119)

ˆ$ and Aˆ$ in Eq. (3.108) are equivalent to those in Here we should note that the expressions of V the HLS with unitary gauge. Then, together with this fact, Eqs. (3.114)–(3.116) and (3.119) show that all the building blocks of the Lagrangian of the matter ?eld method are expressed by the building blocks of the HLS Lagrangian. Therefore, for any Lagrangian of the matter ?eld method consisting of such building blocks, whatever the form it takes, we can construct the equivalent Lagrangian of the HLS. Let us express the Lagrangian in Eq. (3.112) using the building blocks of the HLS, and obtain the relations between the parameters in the matter 0eld method and those in the HLS. The 0rst and ˆ$ and Aˆ$ , so we concentrate on the the third term in Eq. (3.112) are already expressed by #⊥$ , V second term, LC . This is expressed as I2 LC = I2 M2 tr[#ˆ $ #ˆ$ ] − tr[V$ V $ ] 2 + (−3I2 − 12IgV ) tr[#ˆ⊥$ #ˆ$⊥ #ˆ⊥ #ˆ⊥ ] + (−3I2 ) tr[#ˆ $ #ˆ$ #ˆ #ˆ ] + (−2I2 − 4IgV ) tr[#ˆ⊥$ #ˆ⊥ #ˆ$ #ˆ ] + (2I2 + 4IgV ) tr[#ˆ⊥$ #ˆ⊥ #ˆ #ˆ$ ] 2 I + 2IgV (tr[#ˆ⊥$ #ˆ$⊥ ])2 + (I2 + 4IgV ) tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$⊥ #ˆ⊥ ] + 2 2 I (tr[#ˆ $ #ˆ$ ])2 + (I2 ) tr[#ˆ $ #ˆ ] tr[#ˆ$ #ˆ ] + 2

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I2 + − + IfV 2

ˆ$ V ˆ$ V $ ] ˆ $ ] + (I2 + IfV ) tr[V tr[V

+ i(2I2 + 4IgV ) tr[V$ #ˆ$⊥ #ˆ⊥ ] + i(2I2 ) tr[V$ #ˆ$ #ˆ ] ˆ$ #ˆ$ #ˆ⊥ ] + i(2I2 + 2IfV ) tr[V ˆ$ #ˆ$ #ˆ ] ; + i(−2I2 − 2IfV − 4IgV ) tr[V ⊥

(3.120)

where we used the following relation valid for Nf = 3: tr[#ˆ⊥$ #ˆ⊥ #ˆ$⊥ #ˆ⊥ ] = tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$⊥ #ˆ⊥ ] +

1 (tr[#ˆ⊥$ #ˆ$⊥ ])2 − 2 tr[#ˆ⊥$ #ˆ$⊥ #ˆ⊥ #ˆ⊥ ] ; 2

1 (tr[#ˆ $ #ˆ$ ])2 − 2 tr[#ˆ $ #ˆ$ #ˆ #ˆ ] : (3.121) 2 The combination of the above LC with LCCWZ and Pi terms as in Eq. (3.112) gives the Lagrangian of the matter 0eld method written by using the HLS 0elds. To compare this with the HLS Lagrangian we need to include the higher order terms in addition to the terms given in Eq. (3.38). As we will show in Section 4, we can perform the systematic low-energy derivative expansion in the HLS. The O(p4 ) terms in the counting scheme are listed in Eqs. (4.25)–(4.27) in Section 4.3. Thus, the comparison of the LCS written in terms of the HLS 0eld with the HLS Lagrangian including O(p4 ) terms leads to the relations between the parameters in two methods. First, comparing the second and third terms in Eq. (3.38) with the 0rst and second terms in Eq. (3.120), we obtain tr[#ˆ $ #ˆ #ˆ$ #ˆ ] = tr[#ˆ $ #ˆ ] tr[#ˆ$ #ˆ ] +

1 = I2 ; g2

F2 = I2 M2 :

(3.122)

Second, comparing the yi terms of the HLS in (4.25) with the third to tenth terms in Eq. (3.120) combined with Pi (i = 1; 2; 3) terms, we obtain y1 = −3I2 − 12IgV + 16 (C) 3 ; y6 = −2I2 − 4IgV ; y10 =

y3 = −3I2 ;

y7 = I2 + 4IgV ;

I2 + 2IgV + 16 (C) 1 ; 2

y11 = I2 + 4IgV + 16 (C) ; 2

I2 ; y13 = I2 : (3.123) 2 Finally, comparing the zi terms of the HLS in (4.27) with the eleventh to sixteenth terms in Eq. (3.120) combined with Pi (i = 9; 10) terms, we obtain y12 =

z1 = −

I2 − IfV + (C) 10 ; 2

z3 = I2 + IfV ; z5 = 2I2 ;

z2 = − (C) 10 ;

z4 = 2I2 + 4IgV ;

z6 = −2I2 − 2IfV − 4IgV − 8 (C) ; 9

z7 = 2I2 + 2IfV :

(3.124)

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43

Now let us discuss the number of the parameters in two methods. The Lagrangian of the HLS is given by the sum of the O(p2 ) terms in Eq. (4.20) and O(p4 ) terms in Eqs. (4.25)–(4.27), which we call LHLS(2+4) . The Lagrangian of the matter 0eld method include the mass and kinetic terms of the vector meson and the interaction terms with one vector meson 0eld in addition to the O(p2 ) + O(p4 ) terms of the ChPT Lagrangian. Then we consider yi (i = 1, 10, 11) and zi (i = 1; : : : ; 5) terms in addition to the leading order terms in LHLS(2+4) . First of all, LCCWZ in Eq. (3.112) exactly agrees with LA in Eq. (4.20) or Eq. (3.38), so that we consider other terms. For the four-point interaction of the pseudoscalar mesons, LCS as well as LHLS(2+4) include three independent terms: There are (i = 1; 2; 3) in LCS and yi (i = 10; 11; 1) in LHLS(2+4) . Similarly, correspondences between (C) i (C) correspond to z1 − z2 comparing the terms with the external gauge 0elds, we see that (C) 10 and 9 and z6 , respectively. The remaining parameters in LCS are M , fV and gV , while those in LrmHLS(2+4) are F ; g, z3 and z4 . One might think that the HLS Lagrangian contains more parameters than the matter 0eld Lagrangian does. However, one of F ; g; z3 and z4 can be absorbed into rede0nition of the vector meson 0eld [178] as far as we disregard the counting scheme in the HLS and take LHLS(2+4) as just a model Lagrangian. Then the numbers of the parameters in two methods exactly agree with each other as far as the on-shell amplitude is concerned. Here we show how one of F ; g; z3 and z4 can be absorbed into rede0nition of the vector meson 0eld in the HLS [178]: V$ → V$ + (1 − K)#ˆ $ :

(3.125)

This rede0nition leads to [178] ˆ$ + K(1 − K) i[#ˆ $ ; #ˆ ] + (1 − K) i[#ˆ⊥$ ; #ˆ⊥ ] ; V$ → KV$ + (1 − K)V #ˆ $ → K #ˆ $ :

(3.126)

Then the Lagrangian LHLS(2+4) is changed as K2 tr[V$ V $ ] 2g2 K(1 − K) ˆ$ V $ ] + Kz3 tr[V + − g2 2K(1 − K) +i − + Kz4 tr[V$ #ˆ$⊥ #ˆ⊥ ] + · · · ; 2 g

LHLS(2+4) → K 2 F2 tr[#ˆ $ #ˆ$ ] −

(3.127)

where dots stand for the terms irrelevant to the present discussion. Since K is an arbitrary parameter, we choose K =1−

g2 z4 ; 2

(3.128)

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

so that the fourth term in Eq. (3.127) disappear. The rede0nitions of the other parameters such as F → F =K; z3 →

g → gK ;

z3 1 − K + ;··· ; K g2

(3.129)

give the HLS Lagrangian LHLS(2+4) without z4 term. 12 In rewriting LCS into the HLS form there is an arbitrary parameter I as in Eq. (3.114). This I corresponds to the above parameter K for the rede0nition of the vector meson 0eld in the HLS. We 0x I to eliminate z4 in Eq. (3.124): I = −2gV :

(3.130)

Then we have the following correspondences between the parameters in the HLS and those in the matter 0eld method: 1 = 2gV ; F = 2gV M ; z3 = 2gV (2gV − fV ) : (3.131) g In the above discussions, we have shown that LCS is rewritten into LHLS(2+4) and the number of parameters are exactly same in both Lagrangian. Although the on-shell amplitudes are equivalent in two methods, o/-shell structures are di/erent with each other. This is seen in the coupling g and the – mixing strength g . The on-shell g and g are given by g (p2 = m2 )|HLS = g

gV M2 F2 2 2 = g (p = m )| = ; C 2F2 F2

g (p2 = m2 )|HLS = gF2 [1 − g2 z3 ] = g (p2 = m2 )|C = M2 fV ;

(3.132)

where we add (p2 = m2 ) to express the on-shell quantities. In the low energy limit (p2 = 0), on the other hand, they are given by g (p2 = 0)|HLS = g

F2

= g (p2 = 0)|C = 0 ; 2F2

g (p2 = 0)|HLS = gF2 = g (p2 = 0)|C = 0 :

(3.133)

These implies that two methods give diAerent results for the oA-shell amplitude although they are completely equivalent as far as the on-shell tree-level amplitudes are concerned. We should stress here that the rede0nition in Eq. (3.125) is possible only when we omit the counting scheme in the HLS and regard LHLS(2+4) as the model Lagrangian. When we introduce the systematic derivative expansion in the HLS as we will show in Section 4, the HLS gauge coupling constant g is counted as O(p) while other parameters are counted as O(1). Since the rede0nition in Eq. (3.125) mixes O(p2 ) terms with O(p4 ) terms, we cannot make such a rede0nition. Actually, the rede0nition of the parameters in Eq. (3.129) is inconsistent with the counting rule. As a result of the systematic derivative expansion, all the parameters in the HLS are viable. Thus the complete 12

In Ref. [178] instead of z4 term z3 term is eliminated. Here we think that eliminating z4 term is more convenient since z3 term is needed to explain the deviation of the on-shell KSRF I relation from one. [See Eq. (3.73) and analysis in Section 5.]

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45

equivalence is lost in such a case. Of course, we have not known the systematic derivative expansion including the vector meson in the matter 0eld method, 13 so that the discussion of the equivalence itself does not make sense. 3.7.2. Massive Yang–Mills method The “massive Yang–Mills” 0elds [168,169,192] (for reviews, Refs. [77,141]) for vector mesons ( meson and its >avor partners) and axialvector meson A1 (a1 mesons and its >avor partners) were introduced by gauging the chiral symmetry in the nonlinear chiral Lagrangian in the same manner as the external gauge 0elds ( ; W ± ; Z 0 bosons) in Eqs. (2.4) and (2.54) but were interpreted as vector and axialvector mesons instead of the external gauge bosons. Although axialvector mesons as well as vector mesons must be simultaneously introduced in order that the chiral symmetry is preserved by this gauging, the gauged chiral symmetry is explicitly broken anyway in this approach by the mass of these mesons introduced by hand. Hence the “Massive Yang–Mills” 0eld method as it stands does not make sense as a gauge theory. However, it was shown [198] that the same Lagrangian can be regarded as a gauge-0xed form of the generalized HLS (GHLS) Lagrangian [24,17] which is manifestly gauge-invariant under GHLS. In this sense the GHLS and the Massive Yang–Mills 0eld method are equivalent [198,143,89]. The GHLS is a natural extension of the HLS from Hlocal to Glocal (“generalized HLS”) such that the symmetry Gglobal × Hlocal is extended to Gglobal × Glocal [24,17]. By this the axialvector mesons are incorporated together with the vector mesons as the gauge bosons of the GHLS. Let us introduce dynamical variables by extending Eq. (3.1): U = C†L CM CR ;

(3.134)

where these dynamical variables transform as † CL; R → g˜L; R (x) · CL; R · gL; R ;

(3.135)

CM → g˜L (x) · CM · g˜†R (x) ;

(3.136)

with g˜L; R ∈ Glocal = [SU(Nf )L × SU(Nf )R ]local and gL; R ∈ Gglobal = [SU(Nf )L × SU(Nf )R ]global . The covariant derivatives read: D$ CL = 9$ CL − iL$ CL + iCL L$

(L → R) ;

D$ CM = 9$ CM − iL$ CM + iCM R$ ;

(3.137) (3.138)

where we also have introduced the external gauge 0elds, L$ =R$ = V$ ∓ A$ for gauging the Gglobal in addition to the GHLS gauge bosons L$ =R$ = V$ ∓ A$ for Glocal as in Eq. (3.28). There are four lowest derivative terms invariant under (gauged-Gglobal ) × Glocal : L = aLV + bLA + cLM + dL ;   † † † 2 D$ CL · CL + CM D$ CR · CR CM  LV = F2 tr  2i 13

See discussions in Section 4.1.

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

F2 $ + CM (R$ − R $ )C† )2 ] ; tr[(L$ − L M 4   † † † 2 D$ CL · CL − CM D$ CR · CR CM  LA = F2 tr  2i =

F2 $ − CM (R$ − R $ )C† )2 ] ; tr[(L$ − L M 4   † 2 C · C D $ M M  = F2 tr[A$ A$ ] ; LM = F2 tr  2i =

  † † † † 2 C · C − C D C · C C − D C · C D $ L M $ R $ M L R M M  L = F2 tr  2i 2 $ ] = F tr[∇$ U ∇$ U † ] ; $A = F2 tr[A 4

(3.139)

in addition to the kinetic terms of the HLS and the external gauge bosons, where we de0ned “converted” external 0elds: $ = CL L$ C† − i9$ CL · C† = V $ − A $ ; L L L

(3.140)

$ = CR R$ C† − i9$ CR · C† = V $ + A $ ; R R R

(3.141)

$ → which transform exactly in the same way as the GHLS gauge 0elds L$ and R$ , respectively: L † † g˜L L$ g˜L − i9$ g˜L · g˜L (similarly for L ↔ R). Note that L in Eq. (3.139) is actually the gauged nonlinear chiral Lagrangian, the 0rst term of Eq. (2.31). In this GHLS Lagrangian we have two kinds of independent gauge 0elds, one for Glocal (L$ =R$ ) including the vector () and axialvector (A1 ) mesons and the other for (gauged-)Gglobal (L$ =R$ ) including the external gauge 0elds , W ± , Z 0 . This is an outstanding feature of the whole HLS approach, since the basic dynamical variables CL and CR have two independent source charges, Glocal for GHLS (Hlocal for HLS) and Gglobal . These two kinds of independent gauge 0elds are automatically introduced through the covariant derivative. Now, a particularly interesting parameter choice in the Lagrangian Eq. (3.139) is a = b = c = 1 (d = 0), which actually yields a successful phenomenology for axialvector mesons as well as the vector mesons [23,17]: By taking a special gauge CM = 1, the Lagrangian reads $ )2 ] + F 2 tr[(A$ − A $ )2 ] + F 2 tr[A$ A$ ] F2 tr[(V$ − V 2 1 F2 1 2 2 2 $ $ ) ] + 2F tr tr[∇$ U ∇$ U † ] : + A$ − A =F tr[(V$ − V 2 2 4

(3.142)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

47

√ √ The kinetic term of should be normalized by the rescaling (x) → 2(x), F → 2F . Then the Lagrangian 0nally takes the form: 2 F2 1 $ $ )2 ] + 4F 2 tr + tr[∇$ U ∇$ U † ] : 2F2 tr[(V$ − V A$ − A (3.143) 2 4 This is the basis for the successful phenomenology including the axialvector mesons in addition to the vector mesons [23,17]. From this we can reproduce the HLS Lagrangian with Gglobal × Hlocal in the energy region lower than the axialvector meson mass m ¡ p ¡ mA1 . In this region we can ignore the kinetic term of A$ $ = 0, by which we can solve away the A$ and hence the equation of motion for A$ reads: A$ − 12 A 0eld in such a way that the second term of Eq. (3.143) simply yields zero. Since the 0rst and the third terms of Eq. (3.143) are the same as 2LV and LA terms in Eq. (3.38), we indeed get back the HLS Lagrangian Eq. (3.38) with a = 2. (The same argument can apply to the arbitrary choice of the parameters a, b, c, d in Eq. (3.139), which by solving away A1 reproduces the HLS Lagrangian (3.38) with arbitrary a.) On the other hand, by taking another special gauge CM = U , CL = CR = 1, Eq. (3.139) with a = b = c = 1 (d = 0) is shown to coincide with the otherwise unjusti0ed Massive Yang–Mills Lagrangian [198]: L = F2 tr[(V$ − V$ )2 ] + F2 tr[(A$ − A$ )2 ] +

F2 tr[D$ UD$ U † ] ; 4

(3.144)

with D$ U ≡ 9$ U − iL$ U + iUR$ . This takes the same form as the Massive Yang–Mills Lagrangian when the external 0elds L$ and R$ are switched o/, and hence the GHLS and the Massive Yang– Mills are equivalent to each other [198,143,89]. (Reverse arguments were also made, starting with the Massive Yang–Mills Lagrangian and arriving at the GHLS Lagrangian by a “gauge transformation” [128,166], although in the Massive Yang–Mills notion there is no gauge symmetry in the literal sense.) In spite of the same form of the Lagrangian, however, the meaning of the 0elds is quite di/erent: In the absence of the external 0elds, the GHLS 0elds in this gauge-0xing no longer transform as the gauge 0elds in sharp contrast to the Massive Yang–Mills notion. Namely, the GHLS gauge bosons L$ and R$ actually transform as matter 0elds under global G(⊂ Gglobal × Glocal ): L$ → gL L$ gL† , R$ → gR R$ gR† , and hence the mass term does not contradict the gauge invariance in the GHLS case (This is because the mass term in the GHLS model is from the Higgs mechanism Glocal ×Gglobal → G in much the same way as that in the HLS model.) In the presence of the external gauge 0elds, on the other hand, both the external 0elds and the HLS 0elds do transform as gauge bosons under the same (gauged-)G symmetry which is a diagonal sum of the Glocal and (gauged-)Gglobal . The existence of the two kinds of gauge bosons transforming under the same group are due to the two independent source charges of the GHLS model. [Equation (3.144) was also derived within the notion of the Massive Yang–Mills [166], without clear conceptual origin of such two independent gauge 0elds.] To conclude the Massive Yang–Mills approach can be regarded as a gauge-0xed form of the GHLS model and hence equivalent to the HLS model for the energy region m ¡ p ¡ mA1 , after solving away the axialvector meson 0eld.

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3.7.3. Anti-symmetric tensor ?eld method Let us show the equivalence between the anti-symmetric tensor 0eld method (ATFM) and the HLS. (T) (T) In Refs. [79,70] the vector meson 0eld is introduced as an anti-symmetric tensor 0eld V$ =−V$ , which transforms homogeneously under the chiral symmetry: (T) (T) V$ → h(; gR ; gL ) · V$ · h† (; gR ; gL ) :

(3.145)

The transformation property of the 0eld is same as that of the matter 0eld method. Then the covariant derivative acting on the 0eld is de0ned in the same way as in the matter 0eld method: D$(T) ≡ 9$ − i[# $ ; ] ;

(3.146)

where # $ is given in Eq. (3.99). Other building blocks of the Lagrangian are exactly same as that (T) $ and A $ together with the above in the matter 0eld method: The building blocks are V$ , #⊥ $ , V 14 covariant derivative. (T) For constructing the Lagrangian we should note that the 0eld V$ contains six degrees of freedom. (T) To reduce them to three the mass and kinetic terms of V$ must have the following form [70]: 1 M2 (T)$ (T) (T) (T)6 (T) (T)$ Lkin V$ D6 V ] + v tr[V$ V ] : T = − tr[D 2 4

(3.147)

The interaction terms are constructed from the building blocks shown above. An example of the Lagrangian is given by [70] √ FV $ (T) (T) $ LT = F2 tr[#⊥$ #⊥ ] + Lkin (3.148) T + √ tr[V$ V$ ] + i 2GV tr[V$ [#⊥ ; #⊥ ]] : 2 Now let us rewrite the above Lagrangian in terms of the 0elds of the HLS following Ref. [178]. This is done by introducing the HLS gauge 0eld V$ in the unitary gauge as an auxiliary 0eld. The dynamics is not modi0ed by adding the auxiliary 0eld to the Lagrangian: 1 1 1 (T) LT = LT + K2 tr V$ − # $ − D(T) V$ V $ − # $ − D6(T) V (T)6$ ; (3.149) 2 K K (T) where K is an arbitrary parameter. The terms including the derivative of V$ in LT can be removed by performing the partial integral:

1 (T) (T) (T) tr[#ˆ D(T)$ V$ ] ⇒ − tr[(D$ #ˆ − D #ˆ$ )V$ ] + i tr[[#ˆ$ ; #ˆ ]V$ ] 2 =

14

f+$

i i (T) (T) tr[[#ˆ$ ; #ˆ ]V$ ] − tr[[#ˆ$⊥ ; #ˆ⊥ ]V$ ] 2 2 1 1 $ (T) (T) V$ ] + tr[V $ V$ ] ; − tr[V 2 2

(3.150)

$ $ and A $ by u$ = 2#⊥$ , 5$ = −i#$ , The quantities u$ , 5$ , f+$ and f− used in Ref. [70] are related to #⊥$ , #$ , V $ $ $ = 2V and f− = −2A .

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

49

where we used the identity in Eq. (3.118) to obtain the second expression. Substituting Eq. (3.150) into Eq. (3.149), we obtain Mv2 (T) (T)$ tr[V$ V ] 4 √ 1 i $ (T) (T) ] + K tr[[#ˆ ; #ˆ ]V$ ] − i K − 2GV tr[[#ˆ$⊥ ; #ˆ⊥ ]V$ 2 2 FV 1 1 (T) $ V (T) ] + 1 K2 tr[#ˆ$ #ˆ $ ] : K− √ + K tr[V $ V$ tr[V ]− $

2 2 2 2

$ LT = F2 tr[#⊥$ #⊥ ]+

(3.151)

(T) In the above Lagrangian, we can integrate out V$ 0eld. Then, the Lagrangian becomes

1 K2 $ LT = F2 tr[#⊥$ #⊥ ] + K2 tr[#ˆ$ #ˆ $ ] − tr[V$ V $ ] 2 2Mv2 √ √ 2GV ) K(K − 2FV ) K(K − 2 $ ]+i + tr[V$ V tr[V$ #ˆ$⊥ #ˆ⊥ ] + · · · ; 2Mv2 Mv2

(3.152)

where dots stand for the terms irrelevant to the present analysis. Comparing the above Lagrangian with the leading order HLS Lagrangian in Eq. (3.38) or Eq. (4.20) and the zi terms of the HLS in Eq. (4.27), we obtain the following relations: 1 K2 = g2 2Mv2 √ √ K(K − 2FV ) K(K − 2 2GV ) z3 = ; z4 = : 2Mv2 Mv2 F2 =

K2 ; 2

(3.153)

As was discussed for I in Section 3.7.1, the arti0cial coeVcient K is related to the rede0nition of the vector meson 0eld in the HLS [178]. As far as we omit the counting scheme in the HLS and regard LHLS(2+4) as the model Lagrangian, we eliminate z4 term by the rede0nition. Correspondingly, we 0x K to eliminate z4 in Eq. (3.153): √ K = 2 2GV : (3.154) Then we have the following correspondences between the parameters in the HLS and those in the anti-symmetric tensor 0eld method: F2 = 4GV2 ;

1 4GV2 = ; 2 g Mv2

z3 =

2GV (2GV − FV ) : Mv2

(3.155)

With these relations the Lagrangian in the anti-symmetric tensor 0eld method in Eq. (3.148) is equivalent to the leading order terms and z3 and z4 terms in the HLS Lagrangian. We should again note that the above equivalence holds only for the on-shell amplitudes. For the o/-shell amplitudes the equivalence is lost as we discussed for the matter 0led method in Section 3.7.1.

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

3.8. Anomalous processes In QCD with Nf = 3 there exists a non-Abelian anomaly which breaks the chiral symmetry explicitly. In the e/ective chiral Lagrangian this anomaly is appropriately reproduced by introducing the Wess–Zumino action [193,196]. This can be generalized so as to incorporate vector mesons as dynamical gauge bosons of the HLS [74]. In this subsection, following Refs. [74,24], we brie>y review the way of incorporating vector mesons, and then perform analyses on several physical processes focusing whether the vector dominance is satis0ed in the electromagnetic form factors. Here we restrict ourselves to the Gglobal × Hlocal = [U(3)L × U(3)R ]global × [U(3)V ]local model, with Gglobal being fully gauged by the external gauge 0eld L$ and R$ . Since it is convenient to use the language of di/erential forms in the proceeding discussions, we de0ne the following 1-forms: V ≡ V$ d x$ ;

L ≡ L$ d x$ ;

R ≡ R$ d x$ ;

1 1 (9$ U )U −1 d x$ = (dU )U −1 ; L ≡ U −1 dU = U −1 #U : i i Let denote the transformation of Gglobal × Hlocal : #≡

(3.156)

≡ L (AL ) + V (v) + R (AR ) ;

(3.157)

such that CL; R → eiv CL; R e−iAL; R ; V = dv + i[v; V ];

L = dAL + i[AL ; L];

L = dAL + i[AL ; L] :

(3.158)

The essential point of the Wess–Zumino idea [193] is to notice that the anomaly at composite level should coincide with that at quark level. Therefore the e/ective action 5 which describes low energy phenomena must satisfy the same anomalous Ward identity as that in QCD,

1 Nc 2 3 − (L ↔ R) ; (3.159) tr A (dL) − i dL 5[U; L; R] = − 242 M 4 2 where Nc (=3) is the number of colors. Hereafter, we refer Eq. (3.159) as the Wess–Zumino anomaly equation. The so-called Wess–Zumino action, which is a solution to the Wess–Zumino anomaly equation in Eq. (3.159), is given by [193,196]

Nc 5WZ [U; L; R] = tr (#5 ) + (covariantization) ; (3.160) 2402 M 5 where the integral is over a 0ve-dimensional manifold M 5 whose boundary is ordinary Minkowski space M 4 , and (covariantization) denotes the terms containing the external gauge 0elds L and R [127]. The explicit form of the above action is given by [127,74]

5 5WZ [U; L; R] = C tr (# ) − 5Ci tr[L#3 + RL3 ] M5

− 5C

M4

M4

tr[(dLL + L dL)# + (dR R + R dR)L]

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

− 5Ci

+ 5Ci

2

M4

M4

− 5Ci

tr[RU −1 LUL2 − LU RU −1 #2 ]

M4

+ 5Ci

tr[dL dU RU −1 − dR dU −1 LU ]

M4

M4

5 + Ci 2

+ 5C

M4

51

2

tr[(L#) − (RL) ] + 5Ci

M4

tr[L3 # + R3 L]

tr[(dR R + R dR)U −1 LU − (dLL + L dL)U RU −1 ] tr[LU RU −1 L# + RU −1 LU RL]

1 3 −1 3 −1 −1 2 tr R U LU − L U RU + (U RU L) ; 2

(3.161)

where Nc : (3.162) 2402 In the model possessing the HLS Hlocal = [U(3)V ]local the general solution to the Wess–Zumino anomaly equation in Eq. (3.159) is given by [74]

4 Nc ci L i ; (3.163) 5[C†L CR ; V; L; R] = 5WZ [C†L CR ; L; R] + 162 M 4 i=1 C=

where ci are arbitrary constants 15 and Li are gauge invariant 4-forms which conserve parity and charge conjugation but violate the intrinsic parity: 16 L1 = i tr[#ˆ3L #ˆR − #ˆ3R #ˆL ] ;

(3.164)

L2 = i tr[#ˆL #ˆR #ˆL #ˆR ] ;

(3.165)

L3 = tr[FV (#ˆL #ˆR − #ˆR #ˆL )] ;

(3.166)

1 tr[Fˆ L (#ˆL #ˆR − #ˆR #ˆL ) − Fˆ R (#ˆR #ˆL − #ˆL #ˆR )] ; 2 where the gauge covariant building blocks are given by 1 1 ˆ #ˆR ≡ DCR · C†R = #L − V + Rˆ ; #ˆL ≡ DCL · C†L = #L − V + L; i i L4 =

FV ≡ dV − iV 2 ; 15 16

Fˆ L; R = CL; R FL; R C†L; R ;

The normalization of ci here is di/erent by the factor Nc =(162 ) from that in Eq. (7.49) of Ref. [24]. The intrinsic parity of a particle is de0ned to be even if its parity equals (−1)spin , and odd otherwise.

(3.167)

(3.168)

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

with #L; R =

1 dCL; R · C†L; R ; i

FL = dL − iL2 ;

ˆ = CL LC†L ; L

Rˆ = CR RC†R ;

FR = dR − iR2 :

(3.169)

Other possible terms are written in terms of a linear combination of L1 to L4 . 17 We should note that the low energy theorem for anomalous process is automatically satis0ed, since the additional terms other than 5WZ in Eq. (3.163) are gauge invariant and thus do not contribute to the low energy amplitude governed by the anomaly. We now read from the Lagrangian in Eq. (3.163) the VV, V and vertices. These are given by LVV = −

Nc c3 A$6 tr[9$ V 96 V ] 42 F

= g! A$6 9$ ! 96˜ · ˜ + · · · ; Nc (c4 − c3 ) A$6 tr[{9$ V ; 96 V }] 82 F 1 $6 0 = eg! A 96 A 9$ ! + 9$˜ · ˜ + · · · ; 3

(3.170)

L V V = −

LVV = −

(3.171)

Nc (1 − c4 ) A$6 tr[9$ V 96 V ] 42 F

= e2 g A$6 9$ A 96 A 0 + · · · ;

(3.172)

where g! = −

Nc g 2 c3 ; 82 F

(3.173)

g! = −

Nc g (c4 − c3 ) ; 162 F

(3.174)

Nc (1 − c4 ) : 242 F

(3.175)

g = −

The 3 and V3 vertices are given by

3 Nc LV3 = −i 2 3 1 − (c1 − c2 + c4 ) A$6 tr[V$ 9 96 9 ] 3 F 4 = ieg 3 A$6 A$ 9 0 96 + 9 − + · · · ; 17

(3.176)

In the original version in Ref. [74], six terms were included. However, two of them turned out to be charge-conjugation odd and should be omitted [76,123]. This point was corrected in Ref. [24] with resultant four terms in Eqs. (3.164)–(3.167).

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

ρ, ω

ρ, ω ρ, ω

(a)

53

(b)

ω, ρ

(c)

(d)

Fig. 4. E/ective 0 ∗ ∗ vertex: (a) direct 0 interaction ˙ (1 − c4 ); (b) and (c) through 0 ! and 0 0 interactions ˙ (c4 − c3 ); (d) through !0 0 interaction ˙ c3 .

ω

ρ

(a)

(b)

Fig. 5. E/ective !0 ∗ vertex: (a) direct !0 interaction ˙ (c3 − c4 ); (b) through !0 0 interaction ˙ c3 .

π0

ω

(a)

π+

π-

ρ

+

(permutations)

(b)

Fig. 6. E/ective !0 + − vertex: (a) direct !0 + − interaction ˙ (c1 − c2 − c3 ); (b) through ! interaction ˙ c3 .

LV3 = −i

Nc (c1 − c2 − c3 ) A$6 tr[V$ 9 96 9 ] 42 F3

= ig!3 A$6 !$ 9 0 96 + 9 − ; where g 3

3 Nc 1 − (c1 − c2 + c4 ) ; =− 122 F3 4

g!3 = −

3Nc g (c1 − c2 − c3 ) : 162 F3

(3.177)

(3.178) (3.179)

From the above vertices we construct the e/ective vertices for 0 ∗ ∗ , !0 ∗ , !0 + − and The relevant diagrams are shown in Figs. 4–7. The e/ective vertices are given by

∗ 0 + − .

5$ [0 ; ∗ (q1 ; $); ∗ (q1 ; )]

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π0

π+

ρ

π-

(a)

+

(permutations)

+

(permutations)

(b)

ω

ω

ρ

(d) (c)

Fig. 7. E/ective ∗ 0 + − vertex: (a) direct ∗ 0 + − interaction ˙ g 3 ; (b) through ∗ interaction ˙ (c4 − c3 ); (c) through !0 + − interaction ˙ (c1 − c2 − c3 ); (d) through ! interaction ˙ c3 .

Nc c 4 − c3 $#L {D (q12 ) + D (q22 ) + D! (q12 ) + D! (q22 )} =e A q1# q2L {1 − c4 } + 2 12 F 4

c3 2 2 2 2 + {D (q1 )D! (q2 ) + D! (q1 )D (q2 )} ; (3.180) 2

c4 − c3 Nc $#L $ 0 ∗ 2 + c3 D (q ) ; 5 [!(p; $); ; (k; )] = eg 2 A p# kL (3.181) 8 F 2 2

5$ [!(p; $); 0 (q0 ); + (q+ ); − (q− )] = −g

Nc # L A$#L q0 q+ q− [3(c1 − c2 − c3 ) 162 F3

+ 2c3 {D ((q+ + q− )2 ) + D ((q− + q0 )2 ) + D ((q0 + q+ )2 )}] ; 5$ [ ∗ (p; $); 0 (q0 ); + (q+ ); − (q− )] 3 9 Nc # L A q q q =−e (c1 − c2 + c4 ) + (c1 − c2 − c3 )D! (p2 ) $#L 0 + − 1 − 122 F3 4 4 c4 − c3 3 2 + c3 D! (p ) {D ((q+ + q− )2 ) + D ((q− + q0 )2 ) + 4 2

2 + D ((q0 + q+ ) )} ;

(3.182)

(3.183)

where D (q2 ) and D! (q2 ) are meson and ! meson propagators normalized to one in the low-energy limit: D (0) = D! (0) = 1 :

(3.184)

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In this subsection we use the vector meson propagators at leading order: D (q2 ) =

m2 ; m2 − q2

D! (q2 ) =

m2! : m2! − q2

(3.185)

Now let us perform several phenomenological analyses. Below we shall especially focusing whether the vector dominance (VD) is satis0ed in each form factor. Here we summarize the values of the parameters for VD: c3 + c4 =1 ; (a) VD in 0 ∗ : 2 (b) VD in !0 ∗ : c3 = c4 ; (c) VD in 0 ∗ ∗ : c3 = c4 = 1 ; 4 and c3 = c4 : (3.186) 3 When all the above VDs are satis0ed (complete VD), the values of c1 − c2 , c3 and c4 are 0xed: 1 (3.187) (e) complete VD : c3 = c4 = 1 and c1 − c2 = : 3 We 0rst study the decay width of 0 → . When we take q12 = q22 = 0 in the e/ective vertex in Eq. (3.180), terms including c3 and c4 vanish irrespectively of the detailed forms of the and ! propagators. The resultant vertex is identical with the one by the current algebra [32,2,193,3,15]. 18 The predicted [74] decay width is now given by (d) VD in ∗ 0 + − : c1 − c2 + c4 =

5(0 → ) =

#2 m30 : 643 F2

(3.188)

Using the values [91] m0 = (134:9766 ± 0:0006) MeV ;

(3.189)

# = 1=137:03599976 ;

(3.190)

and F in Eq. (3.66) we obtain 5(0 → )|theo = (7:73 ± 0:04) eV :

(3.191)

This excellently agrees with the experimental value estimated from the 0 life time and the branching fraction of 0 → : 5(0 → )|exp = (7:7 ± 0:6) eV :

(3.192)

Second, we study the 0 electromagnetic form factor (0 ∗ form factor). From Eq. (3.180) this form factor is given by c3 + c4 c3 + c 4 2 + F0 (q ) = 1 − (3.193) [D (q2 ) + D! (q2 )] : 2 4 18

We should note that the low energy theorem for → 3 is also intact.

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In the low energy region, by using the explicit forms of and ! propagators, this is approximated as F0 (q2 ) = 1 + 6

q2 + ··· ; m20

where the linear coeVcient 6 is given by

m20 c3 + c4 m20 + 2 : 6= 4 m2 m!

(3.194)

(3.195)

Using the experimental value of this 6 [91] 6|exp = 0:032 ± 0:004 ;

(3.196)

and the values of masses [91] m0 = 134:9766 ± 0:0006 MeV ; m = 771:1 ± 0:9 MeV ; m! = 782:57 ± 0:12 MeV ;

(3.197)

we estimated the value of (c3 + c4 )=2: c 3 + c4 = 1:06 ± 0:13 : (3.198) 2 This implies that the VD (a) in Eq. (3.186) is well satis0ed. Next we calculate the ! → 0 decay width. From the e/ective vertex in Eq. (3.181), the decay width is expressed as 2 3 m! − m20 c3 + c4 2 3# 0 : (3.199) 5(! → ) = g 2 644 F2 2m! Using the values of masses in Eq. (3.197) and the parameters F , g and (c3 + c4 )=2 in Eqs. (3.66), (3.74) and (3.198), we obtain 5(! → 0 ) = 0:85 ± 0:34 MeV :

(3.200)

This agrees with the experimental value 19 5(! → 0 )|exp = 0:73 ± 0:03 MeV :

(3.201)

On the other hand, when we use the above experimental value and the value of g in Eq. (3.74), we obtain c 3 + c4 (3.202) 2 = 0:99 ± 0:16 ; which is consistent with the value in Eq. (3.198). 19

This value is estimated from the ! total decay width and the branching fraction of ! → 0 shown in Ref. [91].

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We further study the ! → 0 $+ $− decay, which is suitable for testing the VD (b) in Eq. (3.186). From the e/ective vertex in Eq. (3.181), this decay width is expressed as

(m! −m )2 2 0 2m q2 − 4m2$ 5(! → # ) $ 5(! → 0 $+ $− ) = 1 + dq2 3 q2 q2 q2 4m2$ ×

q2 1+ 2 m! − m2

2

4m2 q2 − 2 ! 2 2 (m! − m )

3=2 |F!0 (q2 )| ;

(3.203)

where q2 is the intermediate photon momentum and F!0 (q2 ) is the !0 transition form factor. In the HLS this F!0 (q2 ) is given by [18,19] ˜ (q2 ) ; F!0 (q2 ) = −c˜ + (1 + c)D

(3.204)

where c˜ =

c4 − c3 : c3 + c 4

(3.205)

Using the propagator in Eq. (3.185) with the experimental values of masses and the ratio of two decay widths 5(! → 0 $+ $− ) = (1:10 ± 0:27) × 10−3 ; (3.206) 5(! → 0 ) exp we estimate the value of c˜ as c˜ = 0:42 ± 0:56

or

− 7:04 ± 0:56 :

(3.207)

The second solution is clearly excluded by comparing the !0 transition form factor in Eq. (3.204) with experiment (see, e.g., Refs. [18,19]). Since the error is huge in the 0rst solution, the 0rst solution is consistent with the VD (b) in Eq. (3.186). However, the comparison of the form factor itself with experiment prefers non-zero value of c˜ [40,18,19], and thus the VD (b) is violated. Finally, we study the ! → 0 + − decay width to check the validity of the complete VD (e) in Eq. (3.187). By using the e/ective !0 + − vertex in Eq. (3.182), the decay width is expressed as [74]

E+ E− [|˜q− |2 |˜q+ |2 − (˜q+ · ˜q− )2 ]|F!→3 |2 ; (3.208) 5(! → 0 + − ) = where E+ and E− are the energies of + and − in the rest frame of !, ˜q+ and ˜q− are the momenta of them, and F!→3 = −g

Nc [3(c1 − c2 − c3 ) 162 F3

+2c3 {D ((q+ + q− )2 ) + D ((q− + q0 )2 ) + D ((q0 + q+ )2 )}] :

(3.209)

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When we use c1 − c2 = 1 and c3 = 1, 20 we obtain 5(! → 0 + − ) = 6:9 ± 2:2 MeV

(c1 − c2 = 1 and c3 = 1) ;

(3.210)

where the error mainly comes from the error of g in Eq. (3.74). This is consistent with the experimental value [91] 5(! → 0 + − )|exp = 7:52 ± 0:10 MeV :

(3.211)

On the other hand, if we assume that the complete VD (e) in Eq. (3.187) were satis0ed, we would have c1 − c2 − c3 = −2=3 and c3 = 1 [74]. Then we would obtain 5(! → 0 + − ) = 4:4 ± 1:4 MeV

(complete vector dominance) :

(3.212)

Comparing this value with the experimental value in Eq. (3.211), we conclude that the complete VD (e) in Eq. (3.187) is excluded by the experiment [74]. 21 4. Chiral perturbation theory with HLS In this section we review the chiral perturbation in the hidden local symmetry (HLS) at one loop. First we show that, thanks to the gauge invariance of the HLS, we can perform the systematic derivative expansion with including vector mesons in addition to the pseudoscalar Nambu–Goldstone bosons (Section 4.1). Then, we give the O(p2 ) Lagrangian with including the external 0elds in Section 4.2, and then present a complete list of the O(p4 ) terms following Ref. [177] (Section 4.3). Explicit calculation is done by using the background 0eld gauge [177,105]. (The background 0eld gauge is explained in Section 4.4, and the calculation is done in Section 4.6.) Since the e/ect of quadratic divergences is important in the analyses in the next sections (see Sections 5 and 6), we explain meaning of the quadratic divergence in our approach in Section 4.5. We brie>y summarize a role of the quadratic divergence in the phase transition in Section 4.5.1. Then, we show that the chiral symmetry restoration by the mechanism shown in Ref. [104] also takes place even in the ordinary nonlinear sigma model when we include the e/ect of quadratic divergences (Section 4.5.2). We present a way to include the quadratic divergences consistently with the chiral symmetry in Section 4.5.3. The low-energy theorem of the HLS, g =2g F2 [KSRF (I)] [126,163], was shown to be satis0ed at one-loop level in Ref. [103] by using the ordinary quantization procedure in the Landau gauge. Section 4.7 is devoted to show that the low-energy theorem remains intact in the present background 0eld gauge more transparently. From the one-loop corrections calculated in Section 4.6 we will obtain the RGEs in the Wilsonian sense, i.e., including quadratic divergences, in Section 4.8. As was shown in Ref. [177], the relations (matching) between the parameters of the HLS and the O(p4 ) ChPT parameters should be obtained by including one-loop corrections in both theories, 20

This is obtained by requiring the VD (c) in Eq. (3.74) [for c3 = 1] and no direct !0 + − vertex [for c1 − c2 = 1]. After Ref. [74] the experimental value of the ! width was substantially changed (see p. 16 “History plots” of Ref. [91]). Then the experimental value of the partial width 5(! → 0 + − ) becomes smaller than that referred in Ref. [74]. Nevertheless the prediction of the complete vector dominance is still excluded by the new data. 21

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since one-loop corrections from O(p2 ) Lagrangian generate O(p4 ) contributions. In Section 4.9 we show some examples of the relations. Finally in Section 4.10 we study phase structure of the HLS, following Ref. [107]. We note that convenient formulas and Feynman rules used in this section are summarized in Appendices A and B. A complete list of the divergent corrections to the O(p4 ) terms is shown in Appendix D. 4.1. Derivative expansion in the HLS model In the chiral perturbation theory (ChPT) [190,79,80] (see Section 2 for a brief review) the derivative expansion is systematically done by using the fact that the pseudoscalar meson masses are small compared with the chiral symmetry breaking scale , . The chiral symmetry breaking scale is considered as the scale where the derivative expansion breaks down. According to the naive dimensional analysis (NDA) [135] the loop correction (without quadratic divergence) generally appears with the factor p2 : (4F )2

(4.1)

For the consistency with the derivative expansion, the above factor must be smaller than one, which implies that the systematic expansion breaks down around the energy scale of 4F . Then, the chiral symmetry breaking scale , is estimated as [see Eq. (2.12)] , 4F ∼ 1:1 GeV ;

(4.2)

where we used F = 86:4 MeV estimated in the chiral limit [79,81]. Since the meson and its >avor partners are lighter than this scale, one can expect that the derivative expansion with including vector mesons are possible in such a way that the physics in the energy region slightly higher than the vector meson mass scale can well be studied. On the other hand, axialvector mesons (a1 and its >avor partners) should not be included since their masses are larger than , . It was 0rst pointed by Georgi [85,86] that, thanks to the gauge invariance, the HLS makes possible the systematic expansion including the vector meson loops, particularly when the vector meson mass is light. It turns out that such a limit can actually be realized in QCD when the number of massless >avors Nf becomes large as was demonstrated in Refs. [104,106]. Then one can perform the derivative expansion with including the vector mesons under such an extreme condition where the vector meson masses are small, and extrapolate the results to the real world Nf = 3 where the vector meson masses take the experimental values. The 0rst one-loop calculation based on this notion was done in Ref. [103]. There it was shown that the low-energy theorem of the HLS [23,22] holds at one loop. This low-energy theorem was proved to hold at any loop order in Refs. [95,96] (see Section 7). Moreover, a systematic counting scheme in the framework of the HLS was proposed in Ref. [177]. These analyses show that, although the expansion parameter in the real-life QCD is not very small: m2 ∼ 0:5 ; 2, the procedure seems to work in the real world. (See, e.g., a discussion in Refs. [95,96].)

(4.3)

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Now, let us summarize the counting rule of the present analysis. As in the ChPT in Ref. [79,80], the derivative and the external gauge 0elds L$ and R$ are counted as O(p), while the external source 0elds , is counted as O(p2 ) since the VEV of , in Eq. (2.28) is the square of the pseudoscalar meson mass, , ∼ m2 [see Eqs. (2.6), (2.14) and (2.28)]. Then we obtain the following order assignment: 9$ ∼ L$ ∼ R$ ∼ O(p) ; , ∼ O(p2 ) :

(4.4)

The above counting rules are the same as those in the ChPT. Di/erences appear in the counting rules for the vector mesons between the HLS and a version of the ChPT [70] where the vector mesons are introduced by anti-symmetric tensor 0elds (“tensor 0eld method”). [A brief review of “tensor 0eld method” and its relation to the HLS are given in Section 3.7.3.] In the “tensor 0eld method” the vector meson 0elds are counted as O(1). On the other hand, for the consistency of the covariant derivative shown in Eq. (3.28) HLS forces us to assign O(p) to V$ ≡ g$ : V$ = g$ ∼ O(p) :

(4.5)

Another essential di/erence between the counting rule in the HLS and that in the “tensor 0eld method” is in the counting rule for the vector meson mass. In the latter the vector meson mass is counted as O(1). However, as discussed around Eq. (4.3), we are performing the derivative expansion in the HLS by regarding the vector meson as light. Thus, similarly to the square of the pseudoscalar meson mass, we assign O(p2 ) to the square of the vector meson mass: m2 = g2 F2 ∼ O(p2 ) ;

(4.6)

m2

which is contrasted to ∼ O(1) in the “tensor 0eld method”. Since the vector meson mass becomes small in the limit of small HLS gauge coupling, we should assign O(p) to the HLS gauge coupling g, not to F [177]: g ∼ O(p) :

(4.7)

This is the most important part in the counting rules in the HLS. By comparing the order for g in Eq. (4.7) with that for g$ in Eq. (4.5), the $ 0eld should be counted as O(1). Then the kinetic term of the HLS gauge boson is counted as O(p2 ) which is of the same order as the kinetic term of the pseudoscalar meson: 1 (4.8) − tr[$ $ ] ∼ O(p2 ) : 2 We stress that it is the existence of the gauge invariance that makes the above systematic expansion possible [85,86]. To clarify this point, let us consider a Lagrangian including a massive spin-1 0eld as Lorentz vector 0eld, which is invariant under the chiral symmetry. An example is the Lagrangian including the vector meson 0eld as a matter 0eld in the sense of CCWZ [53,48] (“matter 0led method”). [A brief review of this “matter 0eld method” and its relation to the HLS are given in Section 3.7.1.] The kinetic and the mass terms of the vector meson 0eld (C) is given by [see $ Eq. (3.111)] 1 (C)$ (C)$ LC = − tr[(C) ] + M2 tr[(C) ]; (4.9) $ $ 2

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(C) where (C) transforms as [see Eq. (3.105)] $ is de0ned in Eq. (3.107). The vector meson 0eld $ (C) † (C) $ → h(; gR ; gL ) · $ · h (; gR ; gL ) ;

(4.10)

where h(; gR ; gL ) is an element of SU(Nf )V as given in Eq. (3.26). The form of the propagator of the vector meson is given by

p$ p 1 g$ − ; (4.11) p2 − m2 m2 which coincides with the vector meson propagator in the unitary gauge of the HLS (Weinberg’s meson [185]). The longitudinal part (p$ p -part) carries the factor of 1=m2 which may generate quantum corrections proportional to some powers of 1=m2 . Appearance of a factor 1=m2 is a disaster in the loop calculations, particularly when the vector meson mass is light. Namely, the derivative expansion discussed above breaks down. We note that the situation is similar in the “Massive Yang–Mills” approach and the “tensor 0eld method” reviewed in Section 3.7.3. In the HLS, however, the gauge invariance prevent such a 1=m2 factor from appearing. This can be easily seen by the following vector meson propagator in an RC -like gauge 0xing [103]:

p $ p 1 g$ − (1 − #) 2 ; (4.12) p2 − m2 p − #m2 where # is the gauge 0xing parameter. The propagator in Eq. (4.12) is well de0ned in the limit of m → 0 except for the unitary gauge (# = ∞), while the propagator in Eq. (4.11) is ill-de0ned in such a limit. In addition, the gauge invariance guarantees that all the interactions never include a factor of 1=g2 ˙ 1=m2 , while it may exist for the lack of the gauge invariance. Then all the loop corrections are well de0ned even in the limit of m → 0. Thus the HLS gauge invariance is essential to performing the above derivative expansion. This makes the HLS most powerful among various methods (see Section 3.7) for including the vector mesons based on the chiral symmetry. In the above discussion we explained the systematic expansion in the HLS based on the naive dimensional analysis (NDA). Here we re0ne the argument in order to study the large Nf QCD. First, we note that the loop corrections generally have an additional factor Nf in front of the contribution. Then, the general expression for the loop correction in Eq. (4.1) is rewritten as p2 Nf ; (4.13) (4F (0))2 where we used F (0) for expressing the decay constant at the low-energy limit (i.e., on mass–shell of ). Hence when Nf is crucial, we cannot ignore the factor Nf , and the chiral symmetry breaking scale in Eq. (4.2) should be changed to 4F (0) , ; (4.14) Nf √ which yields , 4F (0)= 3 ∼ m for Nf = 3 case. This implies that the systematic expansion for Nf = 3 QCD is valid in the energy region around and less than the meson mass. For large Nf , existence of Nf in the denominator in Eq. (4.14) indicates that , decrease with Nf increased. Furthermore, in the large Nf QCD, as we will study in detail in Section 6, the chiral symmetry is expected to be restored at a certain number of >avor Nfcrit , and F (0) will vanish. One might think that there would be no applicable energy region near the critical point. However, this is not the

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case: In the present analysis, we include the e/ect of quadratic divergences which is necessary for realizing the chiral restoration (see Section 4.5) as well as for matching the HLS with underlying QCD in Nf =3 QCD (see Section 5). The inclusion of the quadratic divergence implies that the loop corrections are given in terms of the bare parameter F () instead of the on-shell decay constant F (0). Then, the scale at which the theory breaks down in Eq. (4.14) is further changed to ,

4F () : Nf

(4.15)

This is somewhat higher than the chiral symmetry breaking scale in Eq. (4.14), F () ¿ F (0) = 86:4 MeV for Nf = 3 (see Section 5.2), and even dramatically higher F ()F (0) → 0 near the phase transition point. One might still think that the above systematic expansion would break down in such a case, since the quadratic divergences from higher loops can in principle contribute to the O(p2 ) terms. However, even when the quadratic divergences are explicitly included, we think that the systematic expansion is still valid in the following sense: The quadratically divergent correction to the O(p2 ) term at nth loop order takes the form of [2 =2, ]n , where , is de0ned in Eq. (4.15). Then, by requiring the cuto/ be smaller than , , 2 =2, ¡ 1, we can perform the systematic expansion even when the e/ect of quadratic divergences are included. It should be noticed that the condition 2 =2, ¡ 1 is essentially the same as the one needed for the derivative expansion being valid up until the energy scale : p2 =2, ¡ 1 for p2 ¡ 2 . Now, the question is whether the requirement , can be satis0ed in some limit of QCD. One possible limit is the large Nc limit of QCD. As is well known, the mesonic loop corrections are suppressed in this limit and tree diagrams give dominant contributions. Actually, in the large Nc limit, F2 (0) scales as Nc and thus it is natural to assume that the bare parameter F2 () has the same scaling property, F2 () ∼ Nc , 22 which implies that , becomes large in the large Nc limit. On the other hand, the meson masses such as the vector meson mass m do not scale, so that we can introduce the which has no large Nc scaling property. Then in the large Nc limit (with 0xed Nf ), the quadratically divergent correction at nth loop order is suppressed by [2 =2, ]n ∼ [1=Nc ]n . As a result, we can perform the loop expansion with quadratic divergences included in the large Nc limit, and extrapolate the results to the real-life QCD as well as to the large Nf QCD. We will give a quantitative argument on this point in Section 5 by determining the value for the bare decay constant F () from QCD through the Wilsonian matching condition, and show that the phenomenological analysis based on the ChPT with HLS can be done in remarkable agreement with the experiments in much the same sense as the phenomenological analysis in the ordinary ChPT is successfully extended to the energy region higher than the pion mass scale, which is logically beyond the validity region of the ChPT. 4.2. O(p2 ) Lagrangian For complete analysis at one-loop, we need to include terms including the external scalar and pseudoscalar source 0elds S and P, as shown in Ref. [177]. In this subsection we present a complete 22

In Section 5 we will derive this scaling property using the Wilsonian matching condition.

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O(p2 ) Lagrangian of the HLS with including the external source 0elds S and P in addition to the lowest derivative Lagrangian (3.38). The external source 0eld ,, which is introduced in the ChPT, transforms linearly under the chiral $ symmetry as in Eq. (2.28), and does not transform at all under the HLS. Since #⊥ as well as # $ transforms as the adjoint representation of the HLS, it is convenient to convert , into a 0eld ,ˆ in the adjoint representation of the HLS for constructing the HLS Lagrangian. This is done by using the “converters” CL and CR as ,ˆ = CL , C†R = 2BCL (S + i P)C†R ;

(4.16)

which transforms homogeneously under the HLS [see Eq. (3.2) for the transformation properties of CL and CR ]: ,ˆ → h(x) · ,ˆ · h† (x) :

(4.17)

Then the lowest order term is given by 1 L, = F,2 tr[,ˆ + ,ˆ† ] : 4

(4.18)

This source is needed to absorb the point-like transformations of the and 0elds [177], as was the case for the , 0eld introduced in the ChPT [79,80]. When we include an explicit chiral symmetry breaking due to the current quark mass, we may introduce it as the vacuum expectation value (VEV) of the external scalar source 0eld:   m1   ..  : S = M =  (4.19) .   m Nf However, in the present paper, we work in the chiral limit, so that we take the VEV to zero S =0. Now, the complete leading order Lagrangian is given by [21,24,177] L(2) = LA + aLV + Lkin (V$ ) + L, = F2 tr[#ˆ⊥$ #ˆ$⊥ ] + F2 tr[#ˆ $ #ˆ$ ] −

1 1 tr[V$ V $ ] + F,2 tr[,ˆ + ,ˆ† ] ; 2 2g 4

(4.20)

where F, in the fourth term is introduced to renormalize the quadratically divergent correction to the fourth term [105]. In the present analysis we introduced this parameter in such a way that the 0eld ,ˆ does not get any renormalization e/ect. We note that this F, agrees with F at tree level. 4.3. O(p4 ) Lagrangian In this subsection we present a complete list of the O(p4 ) Lagrangian, following Ref. [177]. We should note that, as in the ChPT (see Section 2), the one-loop contributions calculated from the O(p2 ) Lagrangian are counted as O(p4 ), and thus the divergences appearing at one loop are renormalized by the coeVcients of the O(p4 ) terms listed below.

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To construct O(p4 ) Lagrangian we need to include 0eld strengths of the external gauge 0elds L$ and R$ in addition to the building blocks appearing in the leading order Lagrangian in Eq. (4.20): L$ = 9$ L − 9 L$ − i[L$ ; L ] ; R$ = 9$ R − 9 R$ − i[R$ ; R ] :

(4.21)

We again convert these into the 0elds which transform as adjoint representations under the HLS: $ ≡ CL L$ C† ; L L

$ ≡ CR R$ C† ; R R

(4.22)

which transform as $ · h† (x); $ → h(x) · L L $ → h(x) · R $ · h† (x) : R

(4.23)

Moreover, it is convenient to introduce the following combinations of the above quantities: $ + L $ ]; $ ≡ 1 [R V 2 $ ≡ 1 [R $ − L $ ] : A 2

(4.24)

A complete list of the O(p4 ) Lagrangian for general Nf was given in Ref. [177]. 23 For general Nf there are 35 O(p4 ) terms compared with 13 terms in the ChPT [L0 ; L1 ; : : : ; L10 ; H1 and H2 terms; see Eq. (2.45)]: L(4)y = y1 tr[#ˆ⊥$ #ˆ$⊥ #ˆ⊥ #ˆ⊥ ] + y2 tr[#ˆ⊥$ #ˆ⊥ #ˆ$⊥ #ˆ⊥ ] + y3 tr[#ˆ $ #ˆ$ #ˆ #ˆ ] + y4 tr[#ˆ $ #ˆ #ˆ$ #ˆ ] + y5 tr[#ˆ⊥$ #ˆ$⊥ #ˆ #ˆ ] + y6 tr[#ˆ⊥$ #ˆ⊥ #ˆ$ #ˆ ] + y7 tr[#ˆ⊥$ #ˆ⊥ #ˆ #ˆ$ ] + y8 { tr[#ˆ⊥$ #ˆ$ #ˆ⊥ #ˆ ] + tr[#ˆ⊥$ #ˆ #ˆ⊥ #ˆ$ ]} + y9 tr[#ˆ⊥$ #ˆ #ˆ$⊥ #ˆ ] + y10 ( tr[#ˆ⊥$ #ˆ$⊥ ])2 + y11 tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$⊥ #ˆ⊥ ] + y12 ( tr[#ˆ $ #ˆ$ ])2 + y13 tr[#ˆ $ #ˆ ] tr[#ˆ$ #ˆ ] + y14 tr[#ˆ⊥$ #ˆ$⊥ ] tr[#ˆ #ˆ ] + y15 tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$ #ˆ ] + y16 ( tr[#ˆ⊥$ #ˆ$ ])2 + y17 tr[#ˆ⊥$ #ˆ ] tr[#ˆ$⊥ #ˆ ] + y18 tr[#ˆ⊥$ #ˆ ] tr[#ˆ$ #ˆ⊥ ] ; 23

(4.25)

We note that there are errors in the divergent corrections to wi in Table 1 of Ref. [177]. In this report we list corrected ones in Table 20 in Appendix D.

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65

F,2 F,2 $ † L(4)w = w1 2 tr[#ˆ⊥$ #ˆ⊥ (,ˆ + ,ˆ )] + w2 2 tr[#ˆ⊥$ #ˆ$⊥ ] tr[,ˆ + ,ˆ† ] F F + w3

F,2 F,2 $ † tr[ # ˆ # ˆ ( , ˆ + , ˆ )] + w tr[#ˆ $ #ˆ$ ] tr[,ˆ + ,ˆ† ] 4

$

2 2 F F

+ w5

F,2 tr[(#ˆ$ #ˆ⊥$ − #ˆ⊥$ #ˆ$ )(,ˆ − ,ˆ† )] F2

+ w6

F,4 F,4 † 2 tr[( , ˆ + , ˆ ) ] + w ( tr[,ˆ + ,ˆ† ])2 7 F4 F4

+ w8

F,4 F,4 † 2 † 2 tr[( , ˆ − , ˆ ) ] + w 9 4 ( tr[,ˆ − ,ˆ ]) ; F4 F

(4.26)

ˆ$ V ˆ$ V $ ] ˆ $ ] + z2 tr[Aˆ$ Aˆ$ ] + z3 tr[V L(4)z = z1 tr[V + i z4 tr[V$ #ˆ$⊥ #ˆ⊥ ] + i z5 tr[V$ #ˆ$ #ˆ ] ˆ$ #ˆ$ #ˆ⊥ ] + i z7 tr[V ˆ$ #ˆ$ #ˆ ] − i z8 tr[Aˆ$ (#ˆ$ #ˆ + #ˆ$ #ˆ⊥ )] ; + i z6 tr[V ⊥ ⊥

where use was made of the equations of motion: i F,2 1 † † , ˆ − , ˆ − tr[ , ˆ − , ˆ ] + O(p4 ) ; D$ #ˆ$⊥ = −i(a − 1)[#ˆ $ ; #ˆ$⊥ ] − 4 F2 Nf

(4.27)

(4.28)

D$ #ˆ$ = O(p4 ) ;

(4.29)

D V $ = g2 f2 #ˆ$ + O(p4 ) ;

(4.30)

and the identities: $ ; D$ #ˆ⊥ − D #ˆ⊥$ = i[#ˆ $ ; #ˆ⊥ ] + i[#ˆ⊥$ ; #ˆ ] + A

(4.31)

$ − V$ ; D$ #ˆ − D #ˆ $ = i[#ˆ $ ; #ˆ ] + i[#ˆ⊥$ ; #ˆ⊥ ] + V

(4.32)

$ being de0ned in Eq. (4.24). $ and V with A We note that for Nf = 3, similarly to relation (2.35) for the ChPT, using the identity: 1 tr[A2 ] tr[B2 ] + ( tr[AB])2 (4.33) 2 valid for any pair of traceless, hermitian 3 × 3 matrices A and B, we have the following relations: tr[ABAB] = −2 tr[A2 B2 ] +

tr[#ˆ⊥$ #ˆ⊥ #ˆ$⊥ #ˆ⊥ ] = −2 tr[#ˆ⊥$ #ˆ$⊥ #ˆ⊥ #ˆ⊥ ] +

1 ( tr[#ˆ⊥$ #ˆ$⊥ ])2 + tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$⊥ #ˆ⊥ ] ; 2

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tr[#ˆ $ #ˆ #ˆ$ #ˆ ] = −2 tr[#ˆ $ #ˆ$ #ˆ #ˆ ] 1 + ( tr[#ˆ $ #ˆ$ ])2 + tr[#ˆ $ #ˆ ] tr[#ˆ$ #ˆ ] ; 2 tr[#ˆ⊥$ #ˆ #ˆ$⊥ #ˆ ] = −2 tr[#ˆ⊥$ #ˆ$⊥ #ˆ #ˆ ] 1 tr[#ˆ⊥$ #ˆ$⊥ ] tr[#ˆ #ˆ ] + tr[#ˆ⊥$ #ˆ ] tr[#ˆ$⊥ #ˆ ] : (4.34) 2 Then there are 32 independent terms in the O(p4 ) Lagrangian of the HLS in contrast to 12 terms in the ChPT Lagrangian [L1 ; : : : ; L10 ; H1 and H2 terms; see Eq. (2.43)]. For Nf = 2, on the other hand, we have the following identity valid for traceless, hermitian 2 × 2 matrices A, B, C and D: 1 1 1 tr[ABCD] = tr[AB] tr[CD] − tr[AC] tr[BD] + tr[AD] tr[BC] : (4.35) 2 2 2 Then, each of y1 - through y9 -terms is rewritten into a combination of y10 - through y18 -terms: 1 tr[#ˆ⊥$ #ˆ⊥ #ˆ$⊥ #ˆ⊥ ] = tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$⊥ #ˆ⊥ ] − ( tr[#ˆ⊥$ #ˆ$⊥ ])2 ; 2 1 tr[#ˆ⊥$ #ˆ$⊥ #ˆ⊥ #ˆ⊥ ] = ( tr[#ˆ⊥$ #ˆ$⊥ ])2 ; 2 1 tr[#ˆ $ #ˆ #ˆ$ #ˆ ] = tr[#ˆ $ #ˆ ] tr[#ˆ$ #ˆ ] − ( tr[#ˆ $ #ˆ$ ])2 ; 2 1 tr[#ˆ $ #ˆ$ #ˆ #ˆ ] = ( tr[#ˆ $ #ˆ$ ])2 ; 2 1 tr[#ˆ⊥$ #ˆ$⊥ #ˆ #ˆ ] = tr[#ˆ⊥$ #ˆ$⊥ ] tr[#ˆ #ˆ ] ; 2 1 1 1 tr[#ˆ⊥$ #ˆ⊥ #ˆ$ #ˆ ] = tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$ #ˆ ] − ( tr[#ˆ⊥$ #ˆ$ ])2 + tr[#ˆ⊥$ #ˆ ] tr[#ˆ$ #ˆ⊥ ] ; 2 2 2 1 1 1 tr[#ˆ⊥$ #ˆ⊥ #ˆ #ˆ$ ] = tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$ #ˆ ] − tr[#ˆ⊥$ #ˆ ] tr[#ˆ$ #ˆ⊥ ] + ( tr[#ˆ⊥$ #ˆ$ ])2 ; 2 2 2 +

tr[#ˆ⊥$ #ˆ$ #ˆ⊥ #ˆ ] + tr[#ˆ⊥$ #ˆ #ˆ⊥ #ˆ$ ] 1 1 1 = ( tr[#ˆ⊥$ #ˆ$ ])2 − tr[#ˆ⊥$ #ˆ⊥ ] tr[#ˆ$ #ˆ ] + tr[#ˆ⊥$ #ˆ ] tr[#ˆ$ #ˆ⊥ ] ; 2 2 2 1 tr[#ˆ⊥$ #ˆ #ˆ$⊥ #ˆ ] = tr[#ˆ⊥$ #ˆ ] tr[#ˆ$⊥ #ˆ ] − tr[#ˆ⊥$ #ˆ$⊥ ] tr[#ˆ #ˆ ] : 2 Furthermore, similarly to relation (2.38) for the ChPT, we have the following relations: 1 tr[#ˆ⊥$ #ˆ$⊥ (,ˆ + ,ˆ† )] = tr[#ˆ⊥$ #ˆ$⊥ ] tr[,ˆ + ,ˆ† ] ; 2 1 tr[#ˆ $ #ˆ$ (,ˆ + ,ˆ† )] = tr[#ˆ $ #ˆ$ ] tr[,ˆ + ,ˆ† ] : 2

(4.36)

(4.37)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

67

Thus, there are 24 independent terms in the O(p4 ) Lagrangian of the HLS in contrast to 10 terms in the ChPT Lagrangian [L1 , L2 , L4 , L6 , L7 , L8 , L9 , L10 , H1 and H2 terms; see Eq. (2.44)]. At 0rst sight, so many proliferated terms look untractable and one might think that the ChPT with HLS would be useless. However, it is not the case: In the above O(p4 ) Lagrangian all the terms in L(4)y generate vertices with at least four legs. In other words, all the yi terms do not contribute to two or three point functions. In the chiral limit , ˆ = ,ˆ† = 0 (no explicit chiral symmetry breaking due to the current quark masses), the terms in L(4)w do not contribute to the Green functions of the vector and axialvector currents. In the terms in L(4)z the z1 ∼ z3 terms contribute to two-point function, while the contributions from z4 ∼ z8 terms are operative only for N (¿ 3)-point function. Thus, as far as we consider the two-point functions of the vector and axialvector current, only z1 , z2 and z3 terms in the entire O(p4 ) Lagrangian contribute. Let us study the correspondence between the parameters in the HLS and the O(p4 ) ChPT parameters at tree level. By using the method used in Section 3.6, the correspondence for Nf = 3 is obtained as [105] 24 1 1 1 1 L1 ⇔ − z4 + y2 + y10 ; tree 32g2 32 32 16 1 1 1 1 − z4 + y2 + y11 ; 2 tree 16g 16 16 16

L2 ⇔

L3 ⇔ − tree

3 3 1 1 + z4 + y1 − y2 ; 2 16g 16 16 8

L4 ⇔

1 w2 ; 4

L5 ⇔

1 w1 ; 4

tree

tree

L6 ⇔ w 7 ; tree

L 7 ⇔ w9 ; tree

L8 ⇔ (w6 + w8 ) ; tree

1 L9 ⇔ tree 4

1 − z3 g2

−

1 (z4 + z6 ) ; 8

L10 ⇔ −

1 1 + (z3 − z2 + z1 ) ; 2 4g 2

H1 ⇔ −

1 1 + (z3 + z2 + z1 ) ; 8g2 4

tree

tree

H2 ⇔ 2(w6 − w8 ) ; tree

(4.38)

24 We note that in Ref. [105] the contributions from z4 to L1 , L2 and L3 are missing, and the sign in front of (z4 + z6 )=8 in L9 was wrong. They are corrected in Eq. (4.38).

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where we took F, = F . It should be noticed that the above relations are valid only at tree level. As discussed in Ref. [177], since one-loop corrections from O(p2 ) Lagrangian L(2) generate O(p4 ) contribution, we have to relate these at one-loop level where 0nite order corrections appear in several relations. We will show, as an example, the inclusion of such 0nite corrections to the relation for L10 in Section 4.9 [see Eq. (4.249)]. 4.4. Background ?eld gauge We adopt the background 0eld gauge to obtain quantum corrections to the parameters. [For calculation in other gauges, see Ref. [103] for the RC -like gauge and Refs. [95,96] for the covariant gauge.] This subsection is for a preparation to calculate the quantum corrections at one loop in proceeding subsections. The background 0eld gauge was used in the ChPT in Refs. [79,80], and was applied to the HLS in Ref. [177]. In this gauge we can easily obtain the vector meson propagator, which is gauge covariant even at o/-shell, from the two-point function. 25 Thus, in the background 0eld gauge, we can easily perform the o/-shell extrapolation of the gauge 0eld. Furthermore, while in the covariant or RC -like gauge we need to consider the point transformation of the pion 0eld in addition to the counter terms included in the Lagrangian to renormalize the divergence appearing only in the o/-shell amplitude of more than two pions (see, e.g., Ref. [10]), 26 we do not need to consider such a transformation separately in the background 0eld gauge: The occurrence of the external source 0eld ,ˆ (especially the terms quadratic in ,) ˆ in the counter terms is related to the point transformation in the covariant or RC -like gauge (see e.g., Ref. [79]). Now, following Ref. [177] we introduce the background 0elds CSL and CSR as CL; R = CZL; R CSL; R ;

(4.39)

where CZL; R denote the quantum 0elds. It is convenient to write CZL = CZS · CZ†P ;

CZR = CZS · CZP ;

CZP = exp[i Za Ta =F ];

CZS = exp[i Za Ta =F ] ;

(4.40)

with Z and Z being the quantum 0elds corresponding to the NG boson and the would-be NG boson . The background 0eld VS$ and the quantum 0eld Z$ of the HLS gauge boson are introduced as V$ = VS$ + gZ$ :

(4.41)

25 Note that, in the RC -like gauge 0xing [103], the propagator obtained from the two-point function by naive resummation is not gauge covariant at o/-shell, since the two-point function at one loop is not gauge covariant at o/-shell due to the existence of non-Abelian vertex. This is well-known in de0ning the electroweak gauge boson propagators in the standard model, which is solved by including a part of the vertex correction into the propagator through so-called pinch technique (see, e.g., Refs. [63–65]). In the background 0eld gauge, on the other hand, the gauge invariance (or covariance) is manifestly kept, so that the resultant two-point function and then the propagator obtained by resumming it are gauge covariant even at o/-shell (see, e.g., Ref. [66]). 26 In the analysis done in Section 7 in the covariant gauge, the point transformation needed in the 0eld renormalization in Eq. (7.40) is expressed by a certain function F i (=) in Eq. (7.37).

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

We use the following notations for the background 0elds including CSL; R : 1 1 A$ ≡ [9$ CSR · CS†R − 9$ CSL · CS†L ] + [CSR R$ CS†R − CSL L$ CS†L ] ; 2i 2 1 1 V$ ≡ [9$ CSR · CS†R + 9$ CSL · CS†L ] + [CSR R$ CS†R + CSL L$ CS†L ] ; 2i 2

69

(4.42)

which correspond to #ˆ⊥$ and #ˆ $ + V$ , respectively. The 0eld strengths of A$ and V$ are de0ned as V$ = 9$ V − 9 V$ − i[V$ ; V ] − i[A$ ; A ] ; A$ = 9$ A − 9 A$ − i[V$ ; A ] − i[A$ ; V ] :

(4.43)

Note that both V$ and A$ do not include any derivatives of the background 0eld CSR and CSL : 1 V$ = [CSR R$ CS†R + CSL L$ CS†L ] ; 2 1 A$ = [CSR R$ CS†R − CSL L$ CS†L ] : (4.44) 2 Then, V$ and A$ correspond to Vˆ$ and Aˆ $ in Eq. (4.24), respectively. In addition, we use ,S for the background 0eld corresponding to ,: ˆ ,S ≡ 2BCSL (S + i P)CS†R :

(4.45)

It should be noticed that the quantum 0elds as well as the background 0elds CSR; L transform homogeneously under the background gauge transformation, while the background gauge 0eld VS$ transforms inhomogeneously: † CSR; L → h(x) · CSR; L · gR; L ;

VS$ → h(x) · VS$ · h† (x) − i9$ h(x) · h† (x) ; Z → h(x) · Z · h† (x) ; Z → h(x) · Z · h† (x) ; Z$ → h(x) · Z$ · h† (x) :

(4.46)

Thus, the expansion of the Lagrangian in terms of the quantum 0eld manifestly keeps the HLS of the background 0eld VS$ [177]. We adopt the background gauge 0xing in ’t Hooft–Feynman gauge: LGF = − tr [(DS $ Z$ + M ) Z 2] ;

(4.47)

where DS $ is the covariant derivative on the background 0eld: DS $ Z = 9$ Z − i[VS $ ; Z ] ;

(4.48)

M = gF

(4.49)

and

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is the vector meson mass parameter which at the loop order should be distinguished from the on-shell mass m de0ned in Eq. (4.217). The Faddeev–Popov ghost term associated with the gauge 0xing (4.47) is S DS $ DS $ + M2 )C] + · · · ; LFP = 2i tr [C(

(4.50)

where dots stand for the interaction terms of the quantum 0elds , Z , Z Z$ and the FP ghosts C S and C. Now, the complete O(p2 ) Lagrangian, L(2) + LGF + LFP , is expanded in terms of the quantum S The terms which do not include the quantum 0elds are nothing but the 0elds, , Z , Z Z and C, C. 2 original O(p ) Lagrangian with the 0elds replaced by the corresponding background 0elds. The terms which are of 0rst order in the quantum 0elds lead to the equations of motions for the background 0elds: i F,2 1 $ $ † † , S − , S DS $ A = −i(a − 1)[V$ − VS$ ; A ] − − tr[ , S − , S ] + O(p4 ) ; (4.51) 4 F2 Nf $ DS $ (V − VS $ ) = O(p4 ) ;

(4.52)

$ DS VS $ = g2 F2 (V − VS $ ) + O(p4 ) ;

(4.53)

which correspond to Eqs. (4.28)–(4.30), respectively. To write down the terms which are of quadratic order in the quantum 0elds in a compact and uni0ed way, let us de0ne the following “connections”: () S 5$; ab ≡ i tr[((2 − a)V$ + aV$ )[Ta ; Tb ]] ;

(4.54)

() S 5$; ab ≡ i tr[(V$ + V$ )[Ta ; Tb ]] ;

(4.55)

√ () 5$; ab ≡ i a tr[A$ [Ta ; Tb ]] ;

(4.56)

√ () 5$; ab ≡ i a tr[A$ [Ta ; Tb ]] ;

(4.57)

(V V )

# L 5$; ab ≡ −2i tr[VS$ [Ta ; Tb ]] g#L :

(4.58) (V V )

Here one might doubt the minus sign in front of 5$ # L compared with 5$(SS) (S = ; ). However, since g#L = −#L for # = 1; 2; 3, the minus sign is a correct one. Correspondingly, we should use an unconventional metric −g#L to change the upper indices to the lower ones: V ) (V V ) 5$ (VL# ;ab ≡ (−g## )5$ab# L : (4.59) #

Further we de0ne the following quantities corresponding to the “mass” part: () Oab ≡−

+

4 − 3a a2 $ $ tr[[A ; Ta ][A$ ; Tb ]] − tr[[V − VS $ ; Ta ][V$ − VS$ ; Tb ]] 2 2 F,2 tr[(,S + ,S† − 2M ){Ta ; Tb }] ; 2F2

(4.60)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

1 a $ $ () Oab ≡ − tr[[V − VS $ ; Ta ][V$ − VS$ ; Tb ]] − tr[[A ; Ta ][A$ ; Tb ]] ; 2 2 √ 1√ $ () ≡ −i a tr[DS $ A$ [Ta ; Tb ]] − a tr[[A$ ; Ta ][V − VS $ ; Tb ]] Oab 2 a √ $ a tr[[V$ − VS$ ; Ta ][A ; Tb ]] ; − 1− 2 () () ≡ Oba ; Oab (V VL )

Oab#

(VL )

Oab

≡ −4i tr[VS #L [Ta ; Tb ]] ; L

≡ −2iagF tr[A [Ta ; Tb ]] ; #

(VL )

(4.61)

(4.62) (4.63)

(V# ) ≡ 2iagF tr[A [Ta ; Tb ]] ; Oab

Oab

71

L ≡ 2igF tr[(V − VS L )[Ta ; Tb ]] ;

(4.64) (4.65) (4.66) (4.67)

# (V# ) ≡ −2igF tr[(V − VS # )[Ta ; Tb ]] ; Oab

(4.68)

M ≡ 2BM ;

(4.69)

where with the quark mass matrix M being de0ned in Eq. (4.19). Here by using the equation of motion () in Eq. (4.51), Oab is rewritten into √ 1√ $ $ () = − a(1 − a) tr[[A$ ; V − VS $ ][Ta ; Tb ]] − a tr[[A$ ; Ta ][V − VS $ ; Tb ]] Oab 2 a √ $ a tr[[V$ − VS$ ; Ta ][A ; Tb ]] − 1− 2 √ a F,2 − tr[(,S − ,S† )[Ta ; Tb ]] : (4.70) 4 F2 To achieve more uni0ed treatment let us introduce the following quantum 0elds: PZ A ≡ (Za ; Za ; Za# ) ;

(4.71)

where the lower and upper indices of PZ should be distinguished as in Eq. (4.59). Thus the metric acting on the indices of PZ is de0ned by   ab    ; 9AB ≡  ab   −g#L ab   ab    ; ab 9AB ≡    # gL ab

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  9AB ≡  



ab

  : 

ab

(4.72)

−g#L ab The tree-level mass matrix is de0ned by   M; a ab   AB ≡   ; M2 ab M   #L 2 −g M ab

(4.73)

where the pseudoscalar meson mass M; a is de0ned by M;2 a ab ≡

F,2 tr[M {Ta ; Tb }] : F2

(4.74)

Here the generator Ta is de0ned in such a way that the above masses are diagonalized when we introduce the explicit chiral symmetry breaking due to the current quark masses. It should be noticed that we work in the chiral limit in this paper, so that we take M = 0;

or

M; a = 0 :

Let us further de0ne   () () 0 5$; ab 5$; ab    () () $ )AB ≡  0  ; (5  5$; ab 5$; ab   (V# VL ) 0 0 5$; ab   (V ) () () Oab Oab Oab L   (V )  () () AB ≡  O  Oab Oab Oab L  ;   (V# VL ) (V# ) (V# ) Oab Oab Oab

(4.75)

(4.76)

(4.77)

and $ )AB : $ )AB ≡ 9AB 9$ + (5 (D

(4.78)

It is convenient to consider the FP ghost contribution separately. For the FP ghost part we de0ne similar quantities: (CC) S 5$; ab ≡ 2i tr[V$ [Ta ; Tb ]] ;

(4.79)

$ )(CC) ≡ ab 9$ + 5(CC) ; (D ab $; ab

(4.80)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

(CC) ≡ ab M 2 : M ab

73

(4.81)

By using the above quantities the terms quadratic in terms of the quantum 0elds in the total Lagrangian are rewritten into

d 4 x [L(2) + LGF + LFP ] 1 =− 2 A;B +i

AB + O AB ]PZ B $ · D $ )AB + M d 4 x PZ A [(D (CC) ]C b ; $ · D $ )(CC) + M d 4 x CS a [(D ab ab

(4.82)

a;b

where $ · D $ )AB ≡ (D $ )(CC) $ · D (D ab

$ )AA (D $ )BA ; (D A $ )(CC) (CC) : ≡ (D ac (D$ )cb

(4.83) (4.84)

c

The Feynman rules obtained from the above Lagrangian relevant to the present analysis are shown in Appendix B. 4.5. Quadratic divergences In the usual phenomenological study in the ChPT of the pseudoscalar mesons [190,79,81] as well as the calculations in the early stage of the ChPT with HLS [103,95,96,177] only the logarithmic divergence was included. As far as the bare theory is not referred to, the quadratic divergence is simply absorbed into rede0nitions of the parameters. In other words, when we take only the logarithmic divergence into account and make the phenomenological analysis in the energy region around the vector meson mass without referring to the underlying theory, the systematic expansion explained in Section 4.1 perfectly works in the idealized world where the vector meson mass is small. Furthermore, according to the phenomenological analysis done so far (e.g. in Refs. [103,177]), the results can be extended to the real world in which the vector meson mass takes the experimental value. However, as was shown in Refs. [104,106,107], the inclusion of the quadratic divergence is essential to studying the phase transition with referring to the bare theory. Moreover, it was shown [105] that inclusion of the quadratic divergence is needed to match the HLS with the underlying QCD even for phenomenological reason. One might think that the systematic expansion breaks down when the e/ect of quadratic divergences is included. However, as we discussed in Section 4.1, the systematic expansion still works as far as we regard the cuto/ is smaller than the scale at which the e/ective 0eld theory breaks down, ¡ , 4F ()= Nf . In this subsection, before starting one-loop calculations in the ChPT with HLS, we explain meaning of the quadratic divergence in our approach. First, we explain “physical meaning” of the quadratic divergence in our approach in Section 4.5.1: In Section 4.5.1.1 we show the role of the quadratic divergence in the phase transition using the Nambu–Jona–Lasinio (NJL) model; in Section 4.5.1.2 we show that the inclusion of quadratic divergence is essential even in the standard model when we

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match it with models beyond standard model; and in Section 4.5.1.3 we review the phase transition in the CP N −1 model in D(6 4) dimensions, in which the power divergence D−2 is responsible for the restoration of the symmetry. Then, in Section 4.5.2 we show that the chiral symmetry restoration by the mechanism shown in Ref. [104] also takes place even in the ordinary nonlinear sigma model when we include the e/ect of quadratic divergences. As is well known the naive momentum cuto/ violates the chiral symmetry. Then, it is important to use a way to include quadratic divergences consistently with the chiral symmetry. We adopt the dimensional regularization and identify the quadratic divergences with the presence of poles of ultraviolet origin at n = 2 [183]:

k$ k dn k 1 2 dn k 2 → ; → − g$ : (4.85) i(2)n −k 2 (4)2 i(2)n [ − k 2 ]2 2(4)2 In Section 4.5.3, we discuss a problem in the naive cuto/ regularization, and show that the above regularization in Eq. (4.85) solves the problem. 4.5.1. Role of quadratic divergences in the phase transition 4.5.1.1. NJL model. For explaining the “physical meaning” of the quadratic divergence in our approach, we 0rst discuss the quadratic divergence in the Nambu–Jona–Lasinio (NJL) model in four dimensions, which actually plays precisely the same role as our quadratic divergence in HLS model in the chiral phase transition. Let us start with the NJL model with the fermion 0eld carrying the color index: G LNJL = S i $ 9$ + [( S )2 + ( S i 5 )2 ] ; (4.86) 2Nc which is invariant under U(1)L × U(1)R rotation. We should note that we consider only the case of the attractive interaction G ¿ 0. It is convenient to introduce auxiliary 0elds ’ ∼ −2(G=Nc ) S and ∼ −2(G=Nc ) S i 5 , and rewrite Eq. (4.86) into Nc 2 (’ + 2 ) − S (’ + i 5 ) : (4.87) LAux = S i $ 9$ − 2G Then the e/ective potential in the 1=Nc -leading approximation is given by 2

d4 k ’ + 2 − k 2 Nc 2 2 + V (’ = = 0) : (4.88) ln V (’; ) = (’ + ) − 2Nc 2G i(2)4 −k 2 The gap equation is derived from the stationary condition of the e/ective potential in Eq. (4.88). By setting = 0 and writing m ≡ ’ for the solution for ’, it is expressed as

d4 k 1 m = 4mG ; (4.89) 4 2 i(2) m − k 2 where m is the dynamical mass of the fermion. The right-hand side of Eq. (4.89) is divergent, so we need to use some regularizations. The 0rst one is the naive cuto/ regularization, which seems easy to understand the physical meaning. When we use the naive cuto/ regularization, Eq. (4.89) becomes 2 + m2 G 2 2 : (4.90) m = m 2 − m ln 4 m2

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The second one is the proper time regularization (heat kernel expansion), in which the integral is regularized via

∞ 1 → dS exp[ − S(m2 − k 2 )] : (4.91) m2 − k 2 2 1= By using this regularization the gap equation (4.89) becomes m=m

G 2 m 5(−1; m2 =2 ) ; 42

(4.92)

where 5(n; A) is the incomplete gamma function de0ned in Eq. (A.36):

∞ d z n −z 5(n; A) ≡ z e : z A

(4.93)

Noting that 5(−1; m2 =2 ) is approximated as [see Eq. (A.40)] 5(−1; m2 =2 )

2 2 − ln ; m2 m2

(4.94)

we can show that Eq. (4.90) is essentially equivalent to Eq. (4.92) for large m. The third one is the dimensional regularization, in which Eq. (4.89) becomes m = 4mG

5(1 − n=2) ; (4)n=2 (m2 )1−n=2

(4.95)

where 5(x) is the gamma function. We note here that 5(1 − n=2) generates pole for n = 2 as well as that for n = 4, which correspond to the quadratic divergence and the logarithmic divergence, respectively in four space–time dimensions. These correspondences are seen as follows: In the dimensional regularization we can separate the pole for n = 2 with that for n = 4 using the identity:

dn k 1 5(1 − n=2) = n 2 2 i(2) m − k (4)n=2 (m2 )1−n=2 = = =

5(2 − n=2) 1 (4)n=2 (m2 )1−n=2 1 − n=2 1 (4)n=2 (m2 )1−n=2 1 (4)n=2 (m2 )1−n=2

5(2 − n=2) [(2 − n=2) − (1 − n=2)] 1 − n=2

5(3 − n=2) − 5(2 − n=2) 1 − n=2

=

5(3 − n=2) 5(3 − n=2) 1 1 − n=2 2 1 − n=2 1 − n=2 (4) (m ) 2 − n=2 (4)n=2 (m2 )1−n=2

=

1 m2 1 1 − + ··· ; 2 4 1 − n=2 (4) 2 − n=2

(4.96)

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where dots stands for the 0nite terms. In the naive cuto/ regularization, on the other hand, the same integral is evaluated as

4 d4 k 1 d k 1 m2 = − i(2)4 m2 − k 2 i(2)4 −k 2 [m2 − k 2 ][ − k 2 ] =

2 m2 − ln 2 + · · · ; (4)2 (4)2

(4.97)

where dots stand for the 0nite terms. Comparing Eq. (4.96) with Eq. (4.97), we see that the 0rst term in Eq. (4.96) corresponds to the quadratic divergence in Eq. (4.97), while the second term in Eq. (4.96), as is well-known, does to the logarithmic divergence: 2 1 → ; (4.98) 1 − n=2 4 1 → ln 2 : (4.99) 2 − n=2 By using this, Eq. (4.95) gives the same gap equation as that in Eq. (4.90) up to the terms of order m2 =2 :

G 2 2 2 m = m 2 − m ln 2 : (4.100) 4 m From the above argument we can conclude that the three regularization methods are equivalent as far as the gap equation is concerned. Namely, the 1=(n − 2) pole has exactly the same meaning as the quadratic divergence in the naive cuto/ regularization. So, after the above replacement, the cuto/ in three regularizations can be understood as the physical cuto/ above which the theory is not applicable. Now, let us study the phase structure of the NJL model. The gap equation of the NJL model in the form given in Eq. (4.100) is rewritten into 1 2 1 1 3 m · 2 ln 2 = −m ; (4.101) − 4 m G Gcr where 1 2 = 2 : (4.102) Gcr 4 From this we easily see that m can be non-zero (symmetry breaking solution) only if 1=G−1=Gcr ¡ 0. It should be noticed that without quadratic divergence the spontaneous symmetry breaking cannot occur, since the bare theory (1=G ¿ 0) is in the symmetric phase. This phase structure can be also seen by studying the sign of the coeVcient of the ’2 term in the e/ective potential Eq. (4.88). By expanding the e/ective potential in Eq. (4.88) around ’ = 0 (we set = 0), we have 2

d4 k ’ − k2 Nc 2 ’ − 2Nc ln V (’; = 0) − V (’ = = 0) = 2G i(2)4 −k 2

1 2 2 M ’ + ··· ; 2 ’

(4.103)

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where M’2 is evaluated as (in all three regularizations) 1 1 2 − M’ = Nc : (4.104) G Gcr It should be noticed that the 0rst term 1=G in the right-hand-side of Eq. (4.104) is a bare mass of ’ and positive, while the second term 1=Gcr = 2 =(42 ) [Eq. (4.102)] arises from the quadratic divergence and can change the sign of M’2 . By using this we can determine the phase as M’2 ¡ 0 → broken phase ; M’2 ¿ 0 → symmetric phase :

(4.105)

Namely, although the bare theory looks as if it were in the symmetric phase, the quantum theory can be in the broken phase due to the quadratic divergence: The phase change is triggered by the quadratic divergence. 4.5.1.2. Standard model. In this subsection we consider the e/ect of quadratic divergence in the standard model (SM), in which there exists a quadratically divergent correction to the Higgs mass parameter. When we make phenomenological analysis within the framework of the SM without referring to the model beyond the SM, we can absorb the e/ect of quadratic divergence into the mass parameter, and the e/ect does not enter the phenomenological analysis. However, as many people are thinking, the SM may not be an ultimate theory describing the real world, and it is just a low-energy e/ective 0eld theory of some underlying theory. In such a case, the bare Higgs mass parameter should be determined from the underlying theory and must be tuned to be canceled with the quadratic divergence of order 2 to yield an observed value (250 GeV)2 , which is however an enormous 0ne-tuning if the cuto/ is very big, say the Planck scale 1019 GeV, (250 GeV)2 =(1019 GeV)2 ∼ 10−33 1. This is a famous naturalness problem. Here we study how the e/ect of quadratic divergence enters into the relation between the bare Higgs mass parameter and the order parameter (on the order of 250 GeV) in the SM, and show that the bare Higgs mass parameter is actually relevant to the phase structure of the SM. To explain the essential point we switch o/ all the gauge interactions since they are small at the weak scale 250 GeV. Furthermore, we switch o/ all the Yukawa couplings except the one related to the top quark mass. Then, the relevant part of the Lagrangian is given by L = Lkinetic − y( S L tR = + h:c:) + 9$ =† 9$ = − M 2 =† = − 6(=† =)2 ; (4.106) where S L = (tSL ; bSL ) is SU(2)L doublet 0eld for left-handed top and bottom quarks, tR is the singlet 0eld for the right-handed top quark, = is the Higgs 0eld, y is the top Yukawa coupling and 6 the Higgs self-coupling. The Lkinetic is the kinetic terms for L and tR : Lkinetic = S L $ i9$ L + tSR $ i9$ tR : (4.107) Note that both L and tR are in the fundamental representation of SU(3)c . Here we adopt the large Nc approximation to calculate the e/ective potential for the Higgs 0eld = with regarding the SM as a cuto/ theory. In this approximation we need to take account of only the top quark loop, and the resultant e/ective potential for the Higgs 0eld is given by Nc 2 2 † 4 (=† =)2 ; V (=) − V (0) = M= = = + 6bare + y ln 2 (4.108) (4)2 bare ybare =† =

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where 2 − M=2 = Mbare

2Nc 2 y 2 (4)2 bare

(4.109)

received a correction of quadratic divergence. Note that we put the subscript “bare” to clarify that the 2 ¿ 0, the sign opposite to the usual Higgs parameters are those of the bare Lagrangian and set Mbare potential (see the footnote below). From the e/ective potential in Eq. (4.108) we can determine the phase as 2 2 Mbare M ¡ ⇒ M=2 ¡ 0 (broken phase) ; 2 y2 cr ybare 2 2 M Mbare ¿ ⇒ M=2 ¿ 0 (symmetric phase) ; (4.110) 2 y2 cr ybare where

M2 y2

cr

=

Nc 2 : 82

(4.111)

This shows that there exists the critical value for the bare Higgs mass parameter which distinguishes the broken phase (SU(2)L × U(1)Y is spontaneously broken into U (1)em ) from the symmetric one. When the SM is applicable all the way up to the Planck scale ∼ 1019 GeV, Eq. (4.109) implies that the bare Higgs mass parameter must be tuned to be canceled with the quadratic divergence of order 2 to yield an observed value of order (250 GeV)2 , which is an enormous 0netuning: (250 GeV)2 =(1019 GeV)2 ∼ 10−33 1. This is a di/erent version of the famous naturalness problem. 27 We should stress that the above phase structure in Eq. (4.110) implies that the quadratic divergence in the SM model has the same physical meaning as the quadratic divergence of the NJL model explained in the previous subsection has. To clarify the physical meaning of the quadratic divergence, let us regard the SM as an e/ective 0eld theory of some more fundamental theory, and consider the matching condition between the SM and the underlying theory. Below we shall adopt the top quark condensate model (top-mode standard model) [146,147] as an example of underlying theory (see, for a review, e.g., Ref. [199]), and show that it is essential to include the quadratic divergence in the e/ective 0eld theory (i.e., the SM) when we match it with the underlying theory (i.e., the top quark condensate model). The top quark condensate model, which was proposed by Miransky et al. [146,147] and by Nambu [152] independently, provides a natural understanding of the heavy top quark mass: The mass of top quark is roughly on the order of weak scale 250 GeV. In this model, the standard Higgs doublet is entirely replaced by a composite one formed by a strongly coupled short range dynamics 27

In the usual explanation of the naturalness problem, the top Yukawa coupling is neglected and the quadratically divergent correction to the Higgs mass parameter is proportional to the Higgs self-coupling in the one-loop approximation. Then, the relation between the bare Higgs mass parameter and the order parameter in Eq. (4.109) is modi0ed appropriately. Note that the sign in front of the quadratic divergence coming from the Higgs self-interaction is plus instead of minus in 2 Eq. (4.109) and we set Mbare ¡ 0 as usual. Note also that, when we switch on the gauge interaction of SU(2)L × U(1)Y , the gauge boson loop generates the quadratically divergent correction to the Higgs mass parameter which has also positive 2 sign (and again Mbare ¡ 0 in contrast to the top Yukawa case).

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(four-fermion interaction) which triggers the top quark condensate. The Higgs boson emerges as tSt bound state and hence is deeply connected with the top quark itself. The model was further developed by the renormalization group method [136,137,27]. For illustration of the essential point, we switch o/ all the gauge interactions, and furthermore, we keep only the four-fermion coupling for the top quark. Then, the relevant Lagrangian is expressed as [146,147,27] LTMSM = Lkinetic + Gt ( S L tR )(tSR L ) ;

(4.112)

where the kinetic terms for L and tR are given in Eq. (4.107). To obtain the gap equation, we adopt the large Nc approximation. Then, as we obtained in the previous subsection, the gap equation is given by 28

Nc 2 2 2 (4.113) mt = mt Gt 2 − mt ln 2 ; 8 mt where mt is the top quark mass. This gap equation shows that the model has two phases distinguished by the value of the four-fermion coupling constant Gt : 1 1 ¡ cr ⇒ broken phase ; Gt Gt 1 1 ¿ cr ⇒ symmetric phase ; Gt Gt

(4.114)

1 Nc 2 cr = Gt 82

(4.115)

where

is given by the quadratic divergence as before. Let us now obtain the e/ective 0eld theory of the above top quark condensate model following Ref. [27]. For this purpose it is convenient to introduce auxiliary 0elds =0 = Gt−1 tSR , and rewrite the Lagrangian in Eq. (4.112) as 1 † Le/ = Lkinetic − ( S L tR =0 + h:c:) − = =0 : (4.116) Gt 0 To obtain the e/ective Lagrangian in the low-energy scale $ in the Wilsonian sense, we integrate out the high energy mode $ ¡ E ¡ . In the large Nc approximation the e/ective Lagrangian at scale $ is obtained as [27] Le/ = Lkinetic − ( S L tR =0 + h:c:) + Z= ($) 9$ =†0 9$ =0 − M02 ($) =†0 =0 − 60 ($)(=†0 =0 )2 ;

(4.117)

where Z= ($) = 28

Nc 2 ln ; (4)2 $2

(4.118)

Extra factor 1=2 in Eq. (4.113) compared with Eq. (4.100) comes from the projection operators (1 ± 5 )=2 of rightand left-handed fermions.

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M02 ($) = 60 ($) =

1 2Nc − (2 − $2 ) ; Gt (4)2

2Nc 2 ln : (4)2 $2

(4.119)

(4.120)

Note that the bare mass term M02 () = (1=Gt ) (¿ 0) has received a quantum correction of the quadratic divergence (¡ 0) in accord with the gap equation (4.113), and the kinetic term of =0 and the quartic coupling 6 have been generated as quantum corrections. By rescaling the Higgs 0eld as 1 =0 = = ; Z=

(4.121)

Lagrangian (4.117) is rewritten as Le/ = Lkinetic − y($)( S L tR = + h:c:) + 9$ =† 9$ = − M 2 ($) =† = − 6($)(=† =)2 ;

(4.122)

where y($) = M 2 ($) = 6($) =

1 ; Z= ($)

M02 ($) ; Z= ($)

60 ($) : Z=2 ($)

(4.123)

Lagrangian (4.122) has the same form as the SM Lagrangian (4.106) with the parameters renormalized at scale $ in the Wilsonian sense (including the quadratic divergence), except that we are not free to renormalize the parameters: By taking $ → , we have M02 () = 1=Gt , Z= () = 0 and 6() = 0 and we get back to the original top-mode Lagrangian in Eqs. (4.112) or (4.116). Then the parameters must satisfy the following matching conditions (“compositeness condition” [27]; see also Refs. [20,92]): 1

→

1

=0 ;

(4.124)

6bare 6($) → 4 =0 ; 4 y ($) $→ ybare

(4.125)

2 Mbare 1 M 2 ($) → = ; 2 2 y ($) $→ ybare Gt

(4.126)

2 y2 ($) $→ ybare

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where, as usual, we identi0ed the parameters renormalized at scale in the Wilsonian sense with the bare parameters. 29 Provided the matching condition for the Higgs mass parameter in Eq. (4.126), we can easily see that the phase structure of the e/ective 0eld theory (SM) shown in Eq. (4.110) completely agrees with that of the underlying theory (top quark condensate model) shown in Eq. (4.114). This shows that the quadratic divergence in the e/ective 0eld theory (SM) has the same physical meaning as the underlying theory (top quark condensate model) has: The e/ect of quadratic divergence can trigger the phase change in the quantum theory. 4.5.1.3. CP N −1 model. Next we review the phase transition in the CP N −1 model in which the power divergence D−2 is responsible for the restoration of the symmetry (see Ref. [24, Chapter 5]). The CP N −1 model in D(6 4) dimensions is a nonlinear sigma model based on the coset space SU(N )=SU(N −1)×U(1). In its popular form the basic 0eld variable is expressed by an N -component scalar 0eld : t

≡ (=1 ; =2 ; : : : ; =N ) ;

=a ∈ C ;

(4.127)

with the constraint † = N=g

(g : coupling constant) :

(4.128)

The Lagrangian is given by L= = D$ † D$ − 6(† − N=g) ;

(4.129)

where the 0eld 6 is a Lagrange multiplier and the U(1) covariant derivative D$ is given by D$ = (9$ − igA$ ) :

(4.130)

The Lagrangian in Eq. (4.129) is clearly invariant under SU(N )global × U(1)local . The U(1)local gauge 0eld A$ has no kinetic term in Eq. (4.129) and is an auxiliary 0eld, which can be eliminated by using the equation of motion for A$ , ↔ ← i †↔ 9$ (f9$ g = f9$ g − f9$ g) : (4.131) A$ = − 2N Then Lagrangian (4.129) is equivalent to ↔ g († 9$ )2 − 6(† − N=g) : (4.132) L= = 9$ † 9$ + 4N In this form it still retains the U(1)local invariance under the transformation (x) = ei’(x) (x). Since has 2N real components and is constrained by one real condition in Eq. (4.128), one might think that the 0eld variable includes 2N − 1 degrees of freedom. But the system actually possesses the 29

Another way to obtain the matching conditions in Eqs. (4.124)–(4.126) is as follows: By rescaling the Higgs 0eld as = = (1=y)=0 , the SM Lagrangian in Eq. (4.106) is expressed as Le/ = Lkinetic − ( S L tR =0 + h:c:) +

1 2 ybare

9$ =†0 9$ =0 −

2 Mbare 6bare =† =0 − 4 (=†0 =0 )2 ; 2 ybare 0 ybare

where we put the subscript “bare” to clarify that the matching must be done for the bare e/ective 0eld theory. Comparing this Lagrangian with the auxiliary 0eld Lagrangian in Eq. (4.116), we obtain the matching conditions in Eqs. (4.124)–(4.126).

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U(1)local gauge invariance and so we can gauge away one further component of , leaving 2N − 2 degrees of freedom which are exactly the dimension of the manifold CP N −1 = SU(N )= SU(N − 1) × U(1). Let us consider the e/ective action for Lagrangian (4.129). In the leading order of the 1=N expansion it is evaluated as

5[; 6] = d D x[D$ † D$ − 6(† − N=g)] + iN TrLn(−D$ D$ − 6) : (4.133) Because of the SU(N ) symmetry, the VEV of can be written in the form √ t (x) = (0; 0; : : : ; N v) : Then the e/ective action (4.133) gives the e/ective potential for v and 6 as

dD k 1 2 V (v; 6) = 6(v − 1=g) + ln(k 2 − 6) : N i(2)D The stationary conditions of this e/ective potential are given by 1 9V = 26v = 0 ; N 9v

1 9V 1 1 dD k 2 =0 : =v − + D N 96 g i(2) 6 − k 2

(4.134)

(4.135)

(4.136) (4.137)

The 0rst condition (4.136) is realized in either of the cases 6 = 0 (v = 0);

case (i) ;

v = 0 (6 = 0);

case (ii) :

(4.138)

The case (i) corresponds to the broken phase of the U(1) and SU(N ) symmetries, and case (ii) does to the unbroken phase. The second stationary condition (4.137) gives relation between 6 and v. By putting 6 = v = 0 in Eq. (4.137), the critical point g = gcr separating the two phases in Eq. (4.138) is determined as

D − 2 1 dD k 1 1 = = : (4.139) i(2)D −k 2 (D=2 − 1)5(D=2) (4)D=2 gcr Substituting Eq. (4.139) into the second stationary condition (4.137), we obtain

1 1 1 dD k 1 2 v − = − − : i(2)D −k 2 6 − k 2 g gcr

(4.140)

We should note that the power divergence in 1=gcr in Eq. (4.139) becomes quadratic divergence in four-dimension (D = 4): 1 2 = gcr (4)2

for D = 4 ;

and that the stationary condition in Eq. (4.140) is rewritten into 2 1 1 6 +6 2 = − v − ln for D = 4 ; 2 (4) 6 g gcr

(4.141)

(4.142)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

or, combined with Eq. (4.136), 1 1 3 − ; v =v g gcr

83

(4.143)

which is compared with Eq. (4.101) in the NJL model up to sign. From Eqs. (4.136) and (4.140) it turns out that cases (i) and (ii) in Eq. (4.138) correspond, respectively, to (i) g ¡ gcr ⇒ v = 0 6 = 0 (ii) g ¿ gcr ⇒ v = 0; 6 = 0

(broken phase of SU(N )) ; (symmetric phase of SU(N )) :

(4.144)

The case (ii) in Eq. (4.144) implies that due to the power divergence in 1=gcr from the dynamics of the CP N −1 model, the quantum theory can be in the symmetric phase of SU(N ), even if the bare theory with 1=g ¿ 0 is written as if it were in the broken phase. 4.5.2. Chiral restoration in the nonlinear chiral Lagrangian 4.5.2.1. Quadratic divergence and phase transition. Here we show that the chiral symmetry restoration actually takes place even in the usual nonlinear chiral Lagrangian when we include quadratic divergences from the loop e/ect [104,107]. The Lagrangian of the nonlinear sigma model associated with SU(Nf )L × SU(Nf )R → SU(Nf )V symmetry breaking is given by [the 0rst term of Eq. (2.31)] 1 L = [F() ]2 tr[∇$ U ∇$ U † ] ; (4.145) 4 where we used F() for the NG boson decay constant in the nonlinear chiral Lagrangian to distinguish it from the one in the HLS. The covariant derivative ∇$ U is de0ned by [see Eq. (2.29)] ∇$ U = 9$ U − iL$ U + iU R$ :

(4.146)

In the chiral perturbation theory (ChPT) [190,79,80] explained in Section 2 the e/ect of quadratic divergences is dropped by using the dimensional regularization. In other words, the e/ect of quadratic divergences is assumed to be subtracted, and thus F() does not get any renormalization e/ects. As far as the bare theory is not referred to, the quadratic divergence is simply absorbed into a rede0nition of F() . As is done in other cases, this treatment is enough and convenient for the usual phenomenological analysis assuming no phase change. However, as we discussed in Section 4.5.1, when we study the phase structure with referring to the bare theory, we have to include the e/ect of quadratic divergences. The e/ect of quadratic divergences is included through the renormalization group equation (RGE) in the Wilsonian sense. Let us calculate the quadratically divergent correction to the pion decay constant and obtain the RGE for [F() ]2 . The 0eld U in Lagrangian (4.145) includes the pion 0eld as U = exp (2i=F() ). Then, the Lagrangian is expanded in terms of 0eld as L = tr[9$ 9$ ] + [F;()bare ]2 tr[A$ A$ ] + tr[[A$ ; ][A$ ; ]] + · · · ;

(4.147)

where F;()bare denotes the bare parameter and the axialvector external 0eld A$ is de0ned by 1 A$ = (R$ − L$ ) : (4.148) 2

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Then the contributions at tree level and one-loop level to the A$ -A two-point function are calculated as (tree)$ TAA = g$ [F;()bare ]2 ; (1−loop)$ TAA = −g$ Nf A0 (0) = −g$ Nf

2 ; (4)2

(4.149)

where the function A0 is de0ned in Eq. (A.1). The renormalization is done by requiring the following is 0nite: [F;()bare ]2 − Nf

2 = (0nite) : (4)2

(4.150)

From this the RGE for [F() ]2 is calculated as $

2Nf 2 d $ : [F() ($)]2 = d$ (4)2

(4.151)

This is readily solved as [F() ($)]2 = [F() ()]2 −

Nf (2 − $2 ) ; (4)2

(4.152)

where the cuto/ is the scale at which the bare theory is de0ned. By taking $ = 0, this is rewritten as 30 [F() (0)]2 = [F() ()]2 − [F(); cr ()]2 ;

(4.153)

where [F(); cr ()]2 =

Nf 2 : (4)2

(4.154)

We here stress that the quadratic divergence in Eq. (4.153) is nothing but the same kind of the quadratic divergences in Eqs. (4.102), (4.104), (4.111) and (4.141). Eq. (4.153) resembles Eqs. (4.101), (4.104), (4.109) and (4.143). The phase is determined by the order parameter, which is given by [F() (0)]2 , the pole residue of . Then [F() (0)]2 corresponds to M’2 in Eq. (4.104) [although the broken phase corresponds to opposite sign], M 2 (0) in Eq. (4.109), or the left-hand side in Eq. (4.142). The 0rst term in the right-hand side (RHS) of Eq. (4.153) ([F() ()]2 ) corresponds to the 0rst term (1=G) of the RHS in Eq. (4.104), the 0rst term (MD2 ()=y2 ()) of the RHS in Eq. (4.109), or the 0rst term (1=g) of the RHS in Eq. (4.142). The second term in Eq. (4.153) does to the second term of the RHS in Eq. (4.104), the second term of the RHS in Eq. (4.109) or the second term of the RHS in Eq. (4.142). Thus, the quadratic divergence [second term in Eq. (4.153)] of the loop can give rise to chiral symmetry restoration F() (0) = 0 [104,107]. Furthermore, we immediately see that there is a critical value for F() () which distinguishes the broken phase from 30

As we will show in Sections 4.10 and 6, the chiral symmetry restoration in the HLS takes place by essentially the same mechanism. There is an extra factor 1=2 in the second term in Eq. (6.106) compared with that in Eq. (4.153). This factor comes from the loop contribution.

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the symmetric one: (i) [F() ()]2 ¿ [F(); cr ()]2 ⇒ [F() (0)]2 ¿ 0

(broken phase) ;

(ii) [F() ()]2 = [F(); cr ()]2 ⇒ [F() (0)]2 = 0 (symmetric phase) :

(4.155)

Although the bare theory looks as if it were in the broken phase (opposite to the NJL model), the quantum theory can actually be in the symmetric phase for certain value of the bare parameter F() (). We also note that Eq. (4.153) takes a form similar to that in the chiral restoration by the pion loop for the 0nite temperature ChPT [82]: Nf 2 (4.156) [F() (T )]2 = [F() (0)]2 − T ; 12 with the replacement → T . Actually, the term is from precisely the same diagrammatic origin as that of our quadratic divergence Eq. (4.154). This point will also be discussed in Section 8. 4.5.2.2. Quadratic divergence in the systematic expansion. Here we discuss the validity of the derivative expansion of ChPT when we include quadratic divergences. As we discussed in Sections 2.2 and 4.1, the derivative expansion in the ChPT is the expansion in terms of Nf p 2 : (4F )2

(4.157)

When we include quadratic divergences, the correction at one-loop is given by Nf 2 =(4F )2 , that at two-loop by [Nf 2 =(4F )2 ]2 , etc. Then the derivative expansion becomes obscure when we include quadratic divergences: There is no longer exact correspondence between the derivative expansion and the loop expansion. Nevertheless, we expect that we can perform the systematic expansion when N f 2 ¡1 : (4F ())2

(4.158)

The above result in Eq. (4.155) is based on the one-loop RGE. Though the condition in Eq. (4.158) is satis0ed in the broken phase (away from the critical point) where we expect that the expansion works well, the expansion becomes less reliable near the critical point since at critical value Nf 2 =[4F(); cr ()]2 = 1 holds. Nevertheless, for Nf = 2 the model is nothing but the O(4) nonlinear sigma model, and it is well known from the lattice analyses (See, for example, Ref. [151], and references cited therein.) that there exists a phase transition (symmetry restoration) for a certain critical value of the hopping parameter which corresponds to [F() ()]2 . This is precisely what we obtained in the above. Thus, we expect that the above result based on the one-loop RGE is reliable at least qualitatively even though a precise value of the F(); cr () might be changed by the higher loop e/ects. It should be emphasized again that the role of quadratic divergence in the chiral Lagrangian is just to decide which phase the theory is in. Once we know the phase, we can simply forget about the quadratic divergence, and then the whole analysis is simply reduced to the ordinary ChPT with only logarithmic divergence so that the systematic expansion is perfect. The same comments also apply to the ChPT with HLS to be discussed later: Once we decide the phase by choosing the bare parameters relevant to the quadratic divergence, we can forget about

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the quadratic divergence as far as we do not make a matching with the QCD (as in Section 5), the situation being reduced precisely back to the ChPT with HLS without quadratic divergence fully discussed in Section 4.1. Then the systematic expansion becomes perfect. 4.5.3. Quadratic divergence in symmetry preserving regularization Let us here discuss a problem which arises for the naive cuto/ regularization when we consider, for example, the Feynman integral for the vector current correlator. (See, for example, Section 6 of Ref. [34].) A main point can be explained by the following Feynman integral:

k $k d4 k I $ ≡ : (4.159) i(2)4 [m2 − k 2 ]2 When we use the naive cuto/, this integral is evaluated as

d4 k k2 g$ I $ = 4 4 i(2) [m2 − k 2 ]2

d4 k d4 k 1 m2 g$ g$ =− + 4 i(2)4 m2 − k 2 4 i(2)4 [m2 − k 2 ]2

1 2 1 2 g$ 2 2 − + m ln( =m ) + (0nite terms) ; = (4)2 4 2

(4.160)

where the 0rst term of the second line generates the 2 -term of the third line and the second term of the second line does the ln(2 =m2 )-term. However, in the dimensional regularization I $ is rewritten into

dn k k2 g$ $ I = n i(2)n [m2 − k 2 ]2

dn k dn k g$ g$ 1 m2 =− + : (4.161) n i(2)n m2 − k 2 n i(2)n [m2 − k 2 ]2 The coeVcient of n = 2 pole in the 0rst term is 1=n = 1=2 instead of 1=4. Then the result after replacement (4.85) is

1 2 1 2 g$ $ 2 2 − + m ln( =m ) + (0nite terms) : (4.162) I = (4)2 2 2 The coeVcients of the quadratic divergences in Eqs. (4.160) and (4.162) are di/erent from each other by a factor 2. The proper time regularization, which has an explicit cuto/ , agrees with the dimensional one but not with the naive cuto/ regularization. Now, when we apply the above results to the calculation of the vector current correlator (see Eq. (70) of Ref. [34]), the result from the cuto/ regularization in Eq. (4.160) violates the Ward–Takahashi identity, while the one from the dimensional regularization in Eq. (4.162) as well as the proper time one is consistent with it.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

87

Thus the following replacement in the dimensional regularization is suitable to identify the quadratic divergence:

k$ k dn k 2 → − g$ : (4.163) i(2)n [ − k 2 ]2 2(4)2 [This can be seen in Eq. (6.5) of Ref. [183] and discussions below Eq. (6.6).] The 1=(n − 2) pole is essentially the same as the naive cuto/ up to a numerical factor. When we use the proper-time regularization (heat kernel expansion), we have explicit cuto/ to be interpreted physically in the naive sense of cuto/ and of course consistent with the invariance, the result being the same as the dimensional regularization with the above replacement. We also note that the same phenomenon is observed (although not for the quadratic divergence) when we calculate the NG boson propagator in the NJL model. In the naive cuto/ regularization we must carefully choose a “correct” routing of the loop momentum in order to get the chiral-invariant result, namely a massless pole for the NG boson. In both the dimensional and the proper-time regularizations the invariant result is automatic. 4.6. Two-point functions at one loop In this subsection, we calculate the contributions to the two-point functions of the background S $, V S$ and VS$ up until O(p4 ). The Lagrangian relevant to two-point functions contains three 0elds, A parameters F2 , a and g at O(p2 ) and three parameters z1 , z2 and z3 at O(p4 ) [see Eqs. (4.20) and (4.27)]: L(2) |,=0 = F2 tr[#ˆ⊥$ #ˆ$⊥ ] + F2 tr[#ˆ $ #ˆ$ ] − ˆ

1 tr[V$ V $ ] ; 2g2

ˆ $ ] + z2 tr[Aˆ$ Aˆ$ ] + z3 tr[V ˆ$ V ˆ$ V $ ] : L(4)z1 ;z2 ;z3 = z1 tr[V

(4.164) (4.165)

The tree-level contribution from L(2) |,=0 is counted as O(p2 ), while the one-loop e/ect calculated ˆ 2 from the O(p ) Lagrangian as well as the tree-level one from z1 , z2 and z3 terms are counted as O(p4 ). The relevant Feynman rules to calculate the one-loop corrections are listed in Appendix B. In the present analysis it is important to include the quadratic divergences to obtain the RGEs in the Wilsonian sense. Since a naive momentum cuto/ violates the chiral symmetry, we need a careful treatment of the quadratic divergences. Thus we adopt the dimensional regularization and identify the quadratic divergences with the presence of poles of ultraviolet origin at n = 2 [183]. As discussed in the previous subsection, this can be done by the following replacement in the Feynman integrals [see Eq. (4.85)]:

k$ k dn k 1 2 2 dn k → ; → − g$ : (4.166) i(2)n −k 2 (4)2 i(2)n [ − k 2 ]2 2(4)2 On the other hand, the logarithmic divergence is identi0ed with the pole at n = 4 [see Eqs. (4.99) and (A.6)]: 1 + 1 → ln 2 ; jS

(4.167)

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where 2 1 − E + ln(4) ; ≡ S 4−n j

(4.168)

with E being the Euler constant. 31 It is convenient to de0ne the following Feynman integrals to calculate the one-loop corrections to the two-point function:

1 dn k 2 A0 (M ) ≡ ; (4.169) n 2 i(2) M − k 2

dn k 1 B0 (p2 ; M1 ; M2 ) ≡ ; (4.170) 2 n 2 i(2) [M1 − k ][M22 − (k − p)2 ]

(2k − p)$ (2k − p) dn k B$ (p; M1 ; M2 ) ≡ : (4.171) i(2)n [M12 − k 2 ][M22 − (k − p)2 ] These are evaluated in Appendix A.1. Here we just show the divergent parts of the above integrals [see Eqs. (A.12), (A.13) and (A.15)]: 2 M2 − ln 2 ; (4)2 (4)2 1 B0 (p2 ; M1 ; M2 )|div = ln 2 ; (4)2 1 B$ (p; M1 ; M2 )|div = −g$ [22 − (M12 + M22 ) ln 2 ] (4)2 A0 (M 2 )|div =

−(g$ p2 − p$ p )

1 ln 2 : 3(4)2

(4.172) (4.173)

(4.174)

S $ -A S . The relevant diagrams Let us start with the one-loop correction to the two-point function A are shown in Fig. 8. By using the Feynman rules given in Appendix B, it immediately follows that the contributions from the diagrams in Fig. 8(a) – (c) are evaluated as (a)$ TA S b (p) S aA

= c;d

√ √ dn k 1 1 $L (− aM f g ) (− aM fcdb gL ) ; T(b)$ (p) cda i(2)n k 2 − M2 −(k − p)2 AS a AS b

31

In Eq. (4.99) we did not include the 0nite part associated with logarithmic divergence. In Eq. (4.167) we determine the 0nite part by evaluating a logarithmically divergent integral in the dimensional regularization and the cuto/ regularization. In the dimensional regularization we have

1 dn k 1 1 2 : = − ln M i(2)n [M 2 − k 2 ]2 (4)2 jS In the cuto/ regularization, on the other hand, the same integral is evaluated as

d4 k 1 1 = [ln 2 − ln M 2 − 1] ; i(2)4 [M 2 − k 2 ]2 (4)2 where we drop O(M 2 =2 ) contributions. Comparing the above two equations, we obtain the replacement in Eq. (4.167).

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

(a)

(b)

89

(c)

S $ -A S . Vertex with a dot (•) implies that the derivatives acting Fig. 8. One-loop corrections to the two-point function A on the quantum 0elds, while that with a circle (◦) implies that no derivatives are included. Feynman rule for each vertex is shown in appendix B.

=

c;d

dn k 1√ 1 1√ $ −i −i a(2k − p) fcda a(−2k + p) fcdb i(2)n 2 M2 − k 2 2

1 ; T(c)$ (p) −(k − p)2 AS a AS b

dn k 1 1 −(1 − a) (facd fbcd + fbcd facd )g$ : = n 2 c i(2) −k 2 ×

(4.175)

d

Then from the de0nitions in Eqs. (4.169)–(4.171) and facd fbcd = Nf ab ;

(4.176)

c;d (a)$ (a)$ these are written as [note that TA S b (p) = TAA (p)ab ] S aA (a)$ (p) = −Nf aM2 g$ B0 (p2 ; M ; 0) ; TAA (b)$ TAA (p) = Nf

a $ B (p; M ; 0) ; 4

(c)$ (p) = Nf (a − 1)g$ A0 (0) : TAA

(4.177)

Then by using Eqs. (4.172)–(4.174), the divergent contributions are given by aM2 ln 2 ; (4)2 a a (b)$ (p)|div = −g$ Nf [22 − M2 ln 2 ] − (g$ p2 − p$ p )Nf ln 2 ; TAA 2 4(4) 12(4)2 (a)$ (p)|div = −g$ Nf TAA

(c)$ (p)|div = g$ Nf TAA

(a − 1) 2 : (4)2

(4.178)

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

(a)

(b)

(c)

(d)

S$ -V S . Fig. 9. One-loop corrections to the two-point function V

S $ -A S two-point function is given by By summing up these parts, the divergent contribution to A Nf $ TAA (p)|div = − [2(2 − a)2 + 3a2 g2 F2 ln 2 ]g$ 4(4)2 −

Nf a ln 2 (g$ p2 − p$ p ) : (4)2 12

(4.179)

These divergences are renormalized by the bare parameters in the Lagrangian. The tree level contribution with the bare parameters is given by (tree)$ (p2 ) = F;2 bare g$ + 2z2; bare (p2 g$ − p$ p ) : TAA

Thus the renormalization is done by requiring the followings are 0nite: Nf F;2 bare − [2(2 − a)2 + 3a2 g2 F2 ln 2 ] = (0nite) ; 4(4)2 Nf a ln 2 = (0nite) : z2; bare − 2(4)2 12

(4.180)

(4.181) (4.182)

S . The relevant diagrams S$ -V Next we calculate one-loop correction to the two-point function V are shown in Fig. 9. By using the Feynman rules given in Appendix B and Feynman integrals in Eqs. (4.169)–(4.171), these are evaluated as (a)$ 2 $ 2 TV S (p) = −Nf M g B0 (p ; M ; M ) ; SV (b)$ TV S (p) = SV

1 Nf B$ (p; M ; M ) ; 8

(c)$ TV S (p) = SV

(2 − a)2 Nf B$ (p; 0; 0) ; 8

(d)$ $ TV S (p) = −(a − 1) Nf g A0 (0) : SV

From Eqs. (4.172)–(4.174) the divergent parts of the above integrals are evaluated as Nf (a)$ $ TV [ − 2aM2 ln 2 ] ; S (p)|div = g SV 2(4)2

Nf Nf 1 1 2 1 2 (b)$ $ 2 − + M ln − (g$ p2 − p$ p ) TVS VS (p)|div = g ln 2 ; 2 2(4) 2 2 2(4)2 12

(4.183)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

ραc p

k

Cc

σc

k-p

a

Vµ

k-p

a

p

Vνb Vµ

p

Vνb Vµ

(b)

ραc

Cc Vνb

(d)

a Vµ (e)

p

k

Vνb

(c)

σc Vµ

Vνb

πc

k-p

a

a Vµ

p

Cd

ρβd

(a)

k-p

a

p

k

ρβd

91

p

k

k-p

a

p

V νb Vµ

p

p

Vνb

πd

σd (f)

k

(g)

Fig. 10. One-loop corrections to the two-point function VS$ -VS .

(c)$ TV S (p)|div SV

=g

$

Nf Nf (2 − a)2 (2 − a)2 2 $ 2 $ ln 2 ; − − (g p − p p ) 2(4)2 2 2(4)2 12

Nf [ − 2(a − 1)2 ] : 2(4)2 S$ -V S two-point function is given by Then the divergent contribution to V Nf $ [(1 + a2 )2 + 3ag2 F2 ln 2 ]g$ TV S (p)|div = − SV 4(4)2 (d)$ $ TV S (p)|div = g SV

Nf 5 − 4a + a2 ln 2 (g$ p2 − p$ p ) : (4)2 24 The above divergences are renormalized by the bare parameters in the tree contribution: −

(tree)$ 2 2 $ 2 $ $ TV S (p ) = F; bare g + 2z1; bare (p g − p p ) : SV

Thus we require the following quantities are 0nite: Nf [(1 + a2 )2 + 3ag2 F2 ln 2 ] = (0nite) ; F;2 bare − 4(4)2 Nf 5 − 4a + a2 ln 2 = (0nite) : z1; bare − 2(4)2 12

(4.184)

(4.185)

(4.186) (4.187) (4.188)

Now, we calculate the one-loop correction to the two-point function VS$ -VS . The relevant diagrams are shown in Fig. 10. These are evaluated as n Nf B$ (p; M ; M ) + 4Nf (g$ p2 − p$ p )B0 (p2 ; M ; M ) ; TV(a)$ S VS (p) = 2 2 $ 2 TV(b)$ S VS (p) = −Nf M g B0 (p ; M ; M ) ; $ TV(c)$ S VS (p) = −Nf B (p; M ; M ) ;

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

$ 2 TV(d)$ S VS (p) = n Nf g A0 (M ) ; $ 2 TV(e)$ S VS (p) = −2Nf g A0 (M ) ;

TV(fS V)$ S (p) =

1 Nf B$ (p; M ; M ) 8

a2 Nf B$ (p; 0; 0) ; (4.189) 8 where n is the dimension of the space–time. Here we need a careful treatment of n, since we identify the quadratic divergence with a pole at n=2. Then, n in front of the quadratic divergence is regarded as 2, while n in front of the logarithmic divergence as 4: In addition to Eqs. (4.172) and (4.174) we have 2 M2 nA0 (M 2 )|div = 2 − 4 ln 2 ; (4.190) (4)2 (4)2 1 [42 − 4(M12 + M22 ) ln 2 ] nB$ (p; M1 ; M2 )|div = −g$ (4)2 TV(g)$ S VS (p) =

4 ln 2 : (4.191) 3(4)2 From Eqs. (4.172)–(4.174), (4.190) and (4.191) the divergent parts of the above contributions in Eq. (4.189) are evaluated as 32 Nf Nf 20 (a)$ ln 2 ; TVV (p)|div = g$ [ − 42 + 8M2 ln 2 ] + (g$ p2 − p$ p ) 2(4)2 2(4)2 3 −(g$ p2 − p$ p )

(b)$ (p)|div = g$ TVV

Nf [ − 2M2 ln 2 ] ; 2(4)2

(c)$ (p)|div = g$ TVV

Nf Nf 2 ln 2 ; [42 − 4M2 ln 2 ] + (g$ p2 − p$ p ) 2(4)2 2(4)2 3

(d)$ TVV (p)|div = g$

Nf [42 − 8M2 ln 2 ] ; 2(4)2

Nf [ − 42 + 4M2 ln 2 ] ; 2(4)2

Nf Nf 1 1 2 1 2 (f )$ $ 2 ln 2 − + M ln − (g$ p2 − p$ p ) TVV (p)|div = g 2 2(4) 2 2 2(4)2 12 2 Nf Nf a 2 a 2 (g)$ $ $ 2 $ ln 2 : − − (g TVV (p)|div = g p − p p ) 2(4)2 2 2(4)2 12 (e)$ TVV (p)|div = g$

32

(4.192)

We should note that when the contributions from (a) and (d) are added before evaluating the integrals, the sum does not include the quadratic divergence. In such a case, we can regard n in front of the sum as 4. Note also that the sum of (c) and (e) does not include the quadratic divergence.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

σc

σc a µ

k-p p

k

a µ

b

Vν

p

πc

k-p p

ρβd

k

b

p

Vν

a µ

k-p p

(b)

(c)

πc

σc

a µ

V νb

(d)

k

p

Vνb

πd

σd

(a)

93

a µ

V νb

(e)

S$ -VS . Fig. 11. One-loop corrections to the two-point function V

Summing up the above contributions, we obtain Nf (1-loop)$ TVV (p)|div = − [(1 + a2 )2 + 3ag2 F2 ln 2 ]g$ 4(4)2 Nf 87 − a2 ln 2 (p2 g$ − p$ p ) : 2(4)2 12 On the other hand, the tree contribution is given by 1 (tree)$ TVV (p2 ) = F;2 bare g$ − 2 (p2 g$ − p$ p ) : gbare +

(4.193)

(4.194)

The 0rst term in Eq. (4.193) which is proportional to g$ is renormalized by F;2 bare by using the requirement in Eq. (4.187). The second term in Eq. (4.193) is renormalized by gbare by requiring Nf 87 − a2 1 ln 2 = (0nite) : − (4.195) 2 2(4)2 12 gbare S$ -VS to determine the renorWe also calculate the one-loop correction to the two-point function V malization of z3 . The relevant diagrams are shown in Fig. 11. These are evaluated as (a)$ (p) = Nf M2 g$ B0 (p2 ; M ; M ) ; TV V (b)$ (p) = TV V

1 Nf B$ (p; M ; M ) ; 8

a(2 − a) Nf B$ (p; 0; 0) ; 8 a (d)$ (p) = Nf g$ A0 (0) ; TV V 2 1 (e)$ (p) = Nf g$ A0 (M2 ) : TV V 2 (c)$ (p) = TV V

(4.196)

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From Eqs. (4.172)–(4.174) the divergent parts of the above contributions are evaluated as Nf (a)$ (p)|div = g$ [2M2 ln 2 ] ; TV V 2(4)2

Nf Nf 1 1 2 1 2 (b)$ $ 2 M ln 2 ; − − (g$ p2 − p$ p ) (p)| = g + ln TV div V 2(4)2 2 2 2(4)2 12 (c)$ TV (p)|div = −g$ V

Nf a(2 − a) 2 Nf a(2 − a) − (g$ p2 − p$ p ) ln 2 ; 2 2(4) 2 2(4)2 12

(d)$ TV (p)|div = g$ V

Nf a2 ; 2(4)2

(e)$ TV (p)|div = g$ V

Nf [2 − 2M2 ln 2 ] : 2(4)2

(1-loop)$ (p)|div = TV V

Nf [(1 + a2 )2 + 3ag2 F2 ln 2 ]g$ 4(4)2

(4.197)

Thus

Nf 1 + 2a − a2 ln 2 (p2 g$ − p$ p ) : 2(4)2 12 The tree contribution is given by −

(tree)$ (p2 ) = −F;2 bare g$ + z3; bare (p2 g$ − p$ p ) : TV V

(4.198)

(4.199)

The 0rst term in Eq. (4.198) which is proportional to g$ is already renormalized by F;2 bare by using the requirement in Eq. (4.187). The second term in Eq. (4.198) is renormalized by z3; bare by requiring Nf 1 + 2a − a2 ln 2 = (0nite) : z3; bare − (4.200) 2(4)2 12 To summarize, Eqs. (4.181), (4.187), (4.195), (4.188), (4.182) and (4.200) are all what we need to renormalize the Lagrangians in Eqs. (4.164) and (4.165). To check the above calculations we calculate the divergent contributions at one loop also by using the heat kernel expansion with the proper time regularization in Appendix D. 4.7. Low-energy theorem at one loop In this subsection, we show that the low-energy theorem of the HLS in Eq. (3.62) is intact at one-loop level in the low-energy limit. It was 0rst shown in Landau gauge without including the quadratic divergences [103]. Here we demonstrate it in the background 0eld gauge including the quadratic divergences. The proof of the low energy theorem at any loop order [95,96] will be shown in Section 7. In the HLS the o/-shell extrapolation of the vector meson 0elds are well de0ned, since they are introduced as a gauge 0eld. Then we can naturally de0ne the – mixing strength and the coupling for the o/-shell . Although these g and g do not have any momentum dependences at

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

95

the leading order O(p2 ), they generally depend on the momenta of and when we include the loop corrections. We write these dependences on the momenta explicitly by g (p2 ) and g (p2 ; q12 ; q22 ), where p is the momentum and q1 and q2 are the pion momenta. By using these, the low-energy theorem in Eq. (3.63) is now expressed as g (p2 = 0) = 2g (p2 = 0; q12 = 0; q22 = 0) F2 (0) ;

(4.201)

where F (0) implies that it is also de0ned at low-energy limit (on-shell of the massless pion). In Ref. [103] the explicit calculation of the one-loop corrections to the – mixing strength and the coupling was performed in the Landau gauge with an ordinary quantization procedure. It was shown that by a suitable renormalization of the 0eld and parameters the low-energy theorem was satis0ed at one-loop level, and that there were no one-loop corrections in the low-energy limit. In the calculation in Ref. [103] the e/ect from quadratic divergences was disregarded. Here we include them in the background 0eld gauge. S and V S In the background 0eld gauge adopted in the present analysis the background 0elds A include the photon 0eld A$ and the background pion 0eld S as S = 1 9 S + ie A [Q; ] A S + ··· ; F F S = eQA − i [9 ; V S ] S + ··· ; 2F2

(4.202)

where F in these expressions should be regarded as F (0) (residue of the pion pole) to identify the 0eld S with the on-shell pion 0eld. On the other hand, the background 0eld VS$ include the background 0eld as VS$ = gS$ ;

(4.203)

where g is renormalized in such a way that the kinetic term of the 0eld S$ is normalized to be S two-point one. 33 The contribution to the – mixing strength is calculated as that to the VS$ -V function. Then the contribution in the low-energy limit is expressed as S g (p2 = 0) = gTV (p2 = 0) ; V S (p2 ) is de0ned by where p is the momentum and the scalar component TV V p$ p $ S (p2 ) ≡ TVV (p) : TV V p2

(4.204) (4.205)

S two-point On the other hand, the correction to the coupling is calculated from the VS$ -V S S S function and V$ -A# -AL three point function. We can easily show that the correction from the three S # - AL three point point function vanishes at low-energy limit as follows: Let 5$#L denote the VS$ -A L function. Then the coupling is proportional to q1# q2 5$#L , where q1 and q2 denote the momenta of two pions. Since the legs # and L of 5$#L are carried by q1 or q2 , q1# q2L 5$#L generally proportional 33

When we use other renormalization scheme for g, the 0nite wave function renormalization constant Z appears in this relation as VS$ = gZ1=2 S$ . Accordingly, g in Eqs. (4.204) and (4.206) is replaced with gZ1=2 . Note that the explicit form of Z depends on the renormalization scheme for g as well as the renormalization scale, but it is irrelevant to the proceeding analysis, since the same factor Z1=2 appears in both Eqs. (4.204) and (4.206).

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

to two of q12 , q22 and q1 ·q2 . Since the loop integral does not generate any massless poles, this implies that q1# q2L 5$#L vanishes in the low-energy limit q12 = q22 = q1 · q2 = 0, and g (p2 = 0; q12 = 0; q22 = 0) = g

S TV (p2 = 0) V

2F2 (0)

:

(4.206)

Combined with Eq. (4.204), Eq. (4.206) leads to g (p2 = 0) = 2g (p2 = 0; q12 = 0; q22 = 0) F2 (0) ;

(4.207)

which is nothing but the low energy theorem in Eq. (4.201). Note that the quadratic divergences S includes the e/ect of quadratic are included in the above discussions: The scalar component TV V divergences [see Eq. (4.198)]. Therefore, both the corrections to the V - mixing strength and the V coupling in the low-energy limit come from only the scalar component of the two-point function S , and thus the low-energy theorem remains intact at one-loop level even including quadratic VS$ -V divergences. 4.8. Renormalization group equations in the Wilsonian sense The RGEs for g and a above the mass scale with including only the logarithmic divergences were given in Ref. [103]. We need RGEs in the Wilsonian sense to study the phase structure. In Ref. [104] the quadratic divergences are further included for this purpose. In this subsection we calculate the RGEs for the parameters F , F (and a ≡ F2 =F2 ), g, z1 , z2 and z3 of the Lagrangians in Eqs. (4.164) and (4.165) from the renormalization conditions derived in Section 4.6. The renormalization conditions for F and F in Eqs. (4.181) and (4.187) lead to the RGEs for F and F as $

Nf dF2 = [3a2 g2 F2 + 2(2 − a)$2 ] ; d$ 2(4)2

(4.208)

$

Nf dF2 [3ag2 F2 + (a2 + 1)$2 ] ; = d$ 2(4)2

(4.209)

where $ is the renormalization scale. Combining these two RGEs we obtain the RGE for a = F2 =F2 as

Nf $2 da 2 (4.210) (a − 1) 3a(a + 1)g − (3a − 1) 2 : $ =− d$ 2(4)2 F We note here that the above RGEs agree with those obtained in Ref. [103] when we neglect the quadratic divergences. From the renormalization condition for g in Eq. (4.195) the RGE for g is calculated as $

Nf 87 − a2 4 dg2 g ; =− 2(4)2 6 d$

(4.211)

which exactly agrees with the RGE obtained in Ref. [103]. It should be noticed that the values g = 0;

a=1

(4.212)

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are the 0xed points of the RGEs for g and a in Eqs. (4.211) and (4.210). These 0xed points were 0rst found through the RGEs without quadratic divergences [103], which actually survive inclusion of the quadratic divergences [104]. The RGEs for z1 , z2 and z3 are calculated from the renormalization conditions in Eqs. (4.188), (4.182) and (4.200) [177,105]: $

Nf 5 − 4a + a2 d z1 = ; d$ (4)2 24

(4.213)

$

Nf a d z2 = ; d$ (4)2 12

(4.214)

$

Nf 1 + 2a − a2 d z3 = : d$ (4)2 12

(4.215)

We note here that the RGE for z1 exactly agrees with that for z2 when a = 1 [a = 1 is also the 0xed point of RGE (4.210)]. Then z1 − z2 = (constant)

(4.216)

is the 0xed point of the above RGEs when a = 1. The mass of is determined by the on-shell condition: m2 = a(m )g2 (m )F2 (m ) :

(4.217)

Below the m scale, decouples and hence F2 runs by the -loop e/ect alone. The quadratically divergent correction to F2 with including only the -loop e/ect is obtained as in Eq. (4.150). From this the resultant RGE for F below the m scale is obtained as in Eq. (4.151): $

2Nf 2 d () [F ($)]2 = $ d$ (4)2

($ ¡ m ) ;

(4.218)

where F() ($) runs by the loop e/ect of alone for $ ¡ m . This is readily solved analytically, and the solution is given by [see Eq. (4.152)] [F() ($)]2 = [F() (m )]2 −

Nf (m2 − $2 ) : (4)2

(4.219)

Unlike the parameters renormalized in a mass independent scheme, the parameter F() ($) ($ ¡ m ) does not smoothly connect to F ($) ($ ¿ m ) at m scale. We need to include an e/ect of 0nite renormalization. The relation between [F() (m )]2 and F2 (m ) based on the matching of the HLS with the ChPT at m scale will be obtained in the next subsection Eq. (4.236)]. Here we use another convenient way to evaluate the dominant contribution: Taking quadratic divergence proportional to a ( contributions speci0c to the HLS) in Eq. (4.179) and replacing by m , we obtain [105] [F() (m )]2 = F2 (m ) +

Nf a(m ) 2 m ; (4)2 2

(4.220)

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which is actually the same relation as that in Eq. (4.236). Combining Eq. (4.219) with Eq. (4.220), we obtain the following relation between [F() ($)]2 for $ ¡ m and F2 (m )

Nf a(m ) () 2 2 2 2 m − $ 1− for $ ¡ m : (4.221) [F ($)] = F (m ) − (4)2 2 Then the on-shell decay constant is expressed as Nf a(m ) m2 : 1 − F2 (0) = [F() (0)]2 = F2 (m ) − (4)2 2

(4.222)

4.9. Matching HLS with ChPT In Section 4.3 we obtained correspondence between the parameters of the HLS and the O(p4 ) ChPT parameters at tree level. However, one-loop corrections from the O(p2 ) Lagrangian L(2) generate O(p4 ) contributions, and then the correct relations should be determined by including the one-loop e/ect as was done in Ref. [177]. [Note that in Ref. [177] e/ects of quadratic divergences are not included.] In this subsection match the axialvector and vector current correlators obtained in the HLS with those in the ChPT at one loop, and obtain the relations among several parameters, by including quadratic divergences. Let us start with the two-point functions of the non-singlet axialvector and vector currents:

a b i d 4 x eipx 0|T J5$ (x)J5 (0)|0 = ab (p$ p − g$ p2 )TA (p2 ) ;

i

d 4 x eiqx 0|T J$a (x)Jb (0)|0 = ab (p$ p − g$ p2 )TV (p2 ) :

(4.223)

In the HLS the axialvector current correlator is expressed as TA(HLS) (p2 ) =

S (p2 ) T⊥ T − T⊥ (p2 ) ; −p2

(4.224)

S T (p2 ) and T⊥ (p2 ) are de0ned from the A$ -A two-point function shown in Section 4.6 where T⊥ by $ S T TAA (p2 ) = g$ T⊥ (p2 ) + (g$ p2 − p$ p )T⊥ (p2 ) :

(4.225)

In the ChPT, on the other hand, the same correlator is expressed as [79,80,177] TA(ChPT) (p2 ) =

(ChPT)S (p2 ) T⊥ + 2Lr10 (m ) − 4H1r (m ) ; −p2

(4.226)

(ChPT)S S is de0ned in a way similar to T⊥ where we set $ = m in the parameters Lr10 and H1r , and T⊥ (ChPT)S 2 in Eq. (4.225). Note that T⊥ (p ) does not depend on the momentum p at one-loop level, then (ChPT)S (ChPT)S T⊥ (p2 ) = T⊥ (0) :

(4.227)

As we stated in Section 3.1 for general case, it is not suitable to extrapolate the form of TA(ChPT) (p2 ) in Eq. (4.226), which is derived at one-loop level, to the energy region around the mass. Instead,

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we take the low-energy limit of TA(HLS) (p2 ), and match TA(HLS) (p2 ) in the low-energy limit with TA(ChPT) (p2 ) in Eq. (4.226). In the HLS for p2 m2 the axialvector current correlator is expressed as [177] 2 S (0) T⊥ p (HLS) 2 S T ; (4.228) − T⊥ (0) − T⊥ (0) + O TA (p ) = 2 −p m2

S (0) is de0ned by where T⊥ d S S 2 T (p ) : T⊥ (0) = ⊥ 2 dp2 p =0

(4.229)

We match the TA(HLS) (p2 ) in Eq. (4.228) with TA(ChPT) (p2 ) in Eq. (4.226) for p2 m2 . We should (ChPT)S S note that we can match the pion pole residue T⊥ (0) with T⊥ (0) separately from the remaining (HLS) 2 terms. It should be noticed that TA (p ) in Eq. (4.228) includes terms higher than O(p4 ) in the counting scheme of the ChPT. (ChPT)S (ChPT)S S Let us 0rst match the T⊥ (0) in Eq. (4.228) with T⊥ (0) = T⊥ (p2 ) in Eq. (4.226). In S 2 the HLS, T⊥ (p ) is calculated as ! 2 M N f S 2(2 − a)$2 − 3aM2 ln 2 + Nf aU (p2 ; M ; 0) ; (p2 ) = F2 ($) − (4.230) T⊥ 4(4)2 $ where U (p2 ; M ; 0) is de0ned by M2 2 U (p ; M ; 0) ≡ {F0 (p2 ; M ; 0) − F0 (0; M ; 0)} (4)2 1 + {FA (p2 ; M ; 0) − FA (0; M ; 0)} 4

;

(4.231)

with the functions F0 and FA given in Appendix A.2. The renormalized F ($) is determined by the following renormalization condition: Nf 1 2 2 2 2 2(2 − a) + 3aM ln − F; bare − 4(4)2 6 Nf [2(2 − a)$2 + 3aM2 ln $2 ] ; (4.232) 4(4)2 where the 0nite part associated with the logarithmic divergence is determined in such a way that the renormalized F without quadratic divergence at $ = M becomes pole residue [103,177]. On the S $ -A S two-point function in the ChPT are calculated other hand, the one-loop corrections to the A from the diagram in Fig. 8(c) with a = 0 taken in the vertex (see also Section 4.5.2). Then, the (ChPT)S T⊥ (p2 ) is given by Nf 2 (ChPT)S T⊥ (p2 ) = [F() ($)]2 − $ ; (4.233) (4)2 where we adopted the following renormalization condition: Nf Nf 2 2 = [F() ($)]2 − $ : (4.234) [F;()bare ]2 − 2 (4) (4)2 = F2 ($) −

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(ChPT)S S It is suitable to match T⊥ (0) with T⊥ (0) with taking $ = m : S T⊥ (0) = F2 (m ) −

Nf (2 − a)m2 2(4)2

(ChPT)S = T⊥ (0) = [F() (m )]2 −

Nf m2 : (4)2

(4.235)

From this we obtain the following parameter relation: [F() (m )]2 = F2 (m ) +

Nf a(m ) 2 m ; (4)2 2

(4.236)

where we also took the renormalization point $=m for a. It should be noticed that this is understood as an e/ect of the 0nite renormalization when we include the e/ect of quadratic divergences. Next we match the non-pole terms in TA(HLS) (p2 ) in Eq. (4.228) with those in TA(ChPT) (p2 ) in T Eq. (4.226). Since z2 ($) does not run for $ ¡ m , the transverse part T⊥ (p2 ) for p2 6 m2 is well approximated by 2z2 (m ), then we have T (0) −2z2 (m ) : − T⊥

(4.237)

S (0) is given by By using the explicit form in Eq. (4.230), T⊥

S − T⊥ (0) =

Nf 11a : (4)2 24

(4.238)

The sum of Eq. (4.237) and Eq. (4.238) should be matched with 2Lr10 (m )−4H1r (m ) in Eq. (4.226). Thus, we obtain 2Lr10 (m ) − 4H1r (m ) = −2z2 (m ) +

Nf 11a(m ) : (4)2 24

(4.239)

S We should note that the second term from T⊥ (0) is the 0nite correction coming from the – loop contribution [105,177]. We further perform the matching for the vector current correlators. The vector current correlator in the HLS is expressed as

TV(HLS) (p2 ) =

TVS (p2 ) [ − TVT (p2 ) − 2TVT (p2 )] − T T (p2 ) ; TVS (p2 ) + p2 TVT (p2 )

(4.240)

where $ TVV (p2 ) = g$ TVS (p2 ) + (g$ p2 − p$ p )T T (p2 ) ; $ (p2 ) = g$ TVS (p2 ) + (g$ p2 − p$ p )TVT (p2 ) ; TVV $ TVV (p2 ) = g$ TVS (p2 ) + (g$ p2 − p$ p )TVT (p2 ) :

(4.241)

Around the mass scale p2 m2 , T T (p2 ), TVT (p2 ) and TVT (p2 ) are dominated by 2z1 (m ), −1=g2 (m ) and z3 (m ), respectively. In the low-energy limit, we need to include the chiral logarithms from the pion loop in addition. These chiral logarithms in T T (p2 ), TVT (p2 ) and TVT (p2 ) are

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evaluated from the diagrams in Figs. 9(c), 10(g) and 11(c), respectively: We should note that the chiral logarithm is included in B$ (p2 ; 0; 0) as [see Eq. (A.10)] B$ (p; 0; 0) = −2g$ A0 (0) − (g$ p2 − p$ p )[B0 (p2 ; 0; 0) − 4B3 (p2 ; 0; 0)]

2 1 8 1 $ $ 2 $ 2 + − ln(−p ) : = −2g − (g p − p p ) (4)2 3(4)2 jS 3

(4.242)

Then, the chiral logarithms are obtained by multiplying (1=3(4)2 ) ln(−p2 =m2 ) by the coeVcients (c)$ (g)$ (c)$ of B$ (p; 0; 0) in TVV (p) in Eq. (4.183), TVV (p) in Eq. (4.189) and TVV (p) in Eq. (4.196), respectively. Noting that 2z1 ($), g($) and z3 ($) do not run for $ ¡ m , we obtain the following approximate forms for p2 m2 : T T (p2 ) 2z2 (m ) + TVT (p2 ) −

(2 − a)2 Nf −p2 ln ; 24 (4)2 m2

a2 Nf 1 −p2 + ln ; g2 (m ) 24 (4)2 m2

TVT (p2 ) z3 (m ) +

a(2 − a) Nf −p2 ln : 24 (4)2 m2

(4.243) (4.244) (4.245)

Thus, for p2 m2 , TV(HLS) (p2 ) in Eq. (4.240) is approximated as [177] TV(HLS) (p2 ) −TVT (p2 ) − 2TVT (p2 ) − T T (p2 )

1 1 Nf −p2 − 2z (m ) − 2z (m ) − ln : 3 1 g2 (m ) 6 (4)2 m2

In the ChPT at one loop the same correlator is expressed as [79,80,177]

−p2 5 1 Nf (ChPT) 2 r r ; ln 2 − (p ) = −2L10 (m ) − 4H1 (m ) − TV 6 (4)2 m 3

(4.246)

(4.247)

where the last term is the 0nite correction. 34 It should be noticed that the coeVcient of the chiral logarithm ln(−p2 =m2 ) in Eq. (4.246) exactly agrees with that in Eq. (4.247). Then, matching Eq. (4.247) with Eq. (4.246), we obtain − 2Lr10 (m ) − 4H1r (m ) +

5Nf 1 = 2 − 2z3 (m ) − 2z1 (m ) : 2 18(4) g (m )

(4.248)

Finally, combining Eq. (4.248) with Eq. (4.239), we obtain the following relation: Lr10 (m ) = −

34

1 4g2 (m

)

+

Nf 11a(m ) Nf 5 z3 (m ) − z2 (m ) + z1 (m ) + + : 2 2 (4) 96 (4)2 72

This 0nite correction was not included in Ref. [105].

(4.249)

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4.10. Phase structure of the HLS In this subsection, following Ref. [107], we study the phase structure of the HLS using the RGEs for F , a and g derived in Section 4.8 [see Eqs. (4.208), (4.210) and (4.211)]. As we demonstrated for the Lagrangian of the nonlinear sigma model in Section 4.5.2, even if the bare Lagrangian is written as if in the broken phase, the quantum theory can be in the symmetric phase. As shown in Eq. (4.155), the phase of the quantum theory is determined from the on-shell decay constant [F() (0)]2 (order parameter): (i) [F() (0)]2 ¿ 0 ⇔ broken phase; (ii) [F() (0)]2 = 0 ⇔ symmetric phase :

(4.250)

In the HLS, we can determine the phase from the order parameter F2 (0) in a similar manner: (i) F2 (0) ¿ 0 ⇔ broken phase ; (ii) F2 (0) = 0 ⇔ symmetric phase :

(4.251)

Before going into the detailed study of the phase structure of the HLS, we here demonstrate, by taking g = 0 and a = 1, 35 that the phase change similar to that in the nonlinear sigma model actually takes place in the HLS [104]. Since the value g = 0 is the 0xed point of the RGE for g in Eq. (4.211) and a = 1 is the one for a in Eq. (4.210) [104], the RGE for F2 in Eq. (4.208) becomes $

Nf 2 d 2 F ($) = $ : d$ (4)2

(4.252)

Since m =0 for g=0, the RGE (4.252) is valid all the way down to the low energy limit, $ ¿ m =0. We should note that there is an extra factor 1/2 in the right-hand side of the RGE (4.252) compared with the RGE (4.151) in the nonlinear sigma model. This is because the (longitudinal ) is the real NG boson in the limit of (g; a) = (0; 1) and it does contribute even for (g; a) = (0; 1). Solution of the RGE (4.252) is given by F2 (0) = F2 () − [Fcr ]2 ; [Fcr ]2 ≡

Nf 2 : 2(4)2

(4.253)

We should note again the extra factor 1/2 compared with Eq. (4.154). As in the case for Eq. (4.153), Eq. (4.253) implies that even if the bare theory of the HLS is written as if it were in the broken phase (F2 () ¿ 0), the quantum theory is actually in the symmetric phase, when we tune the bare parameter as: F2 () = [Fcr ]2 . We should stress that this can occur only if we use the Wilsonian RGEs, i.e., the RGEs including the quadratic divergences. 36 35 As we shall discuss later (see Section 6.1.5), the point (g; a) ≡ (0; 1) should be regarded only as a limit g → 0; a → 1, where the essential feature of the arguments below still remains intact. 36 In the case of large Nf QCD to be discussed in Section 6.3, we shall determine the bare parameter F2 () by the underlying theory through the Wilsonian matching (see Section 5) and hence F2 () is no longer an adjustable parameter, whereas the value of [Fcr ]2 instead of F2 () can be tuned by adjusting Nf .

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For studying the phase structure of the HLS through the RGEs it is convenient to use the following quantities: Nf $2 ; 2(4)2 F2 ($) Nf 2 g ($) ($ ¿ m ) : G($) ≡ 2(4)2 X ($) ≡

(4.254) (4.255)

By using these X ($) and G($), the RGEs in Eqs. (4.208), (4.210) and (4.211) are rewritten as dX = (2 − 3a2 G)X − 2(2 − a)X 2 ; $ (4.256) d$ da = −(a − 1)[3a(a + 1)G − (3a − 1)X ] ; (4.257) $ d$ dG 87 − a2 2 $ (4.258) =− G : d$ 6 As we stated in Section 4.8, the above RGEs are valid above the mass scale m , where m is de0ned by the on-shell condition in Eq. (4.217). In terms of X , a and G, the on-shell condition becomes a(m )G(m ) = X (m ) :

(4.259)

Then the region where the RGEs in Eqs. (4.256)–(4.258) are valid is speci0ed by the condition a($)G($) 6 X ($). We 0rst obtain the 0xed points of the RGEs in Eqs. (4.256)–(4.258). This is done by seeking the parameters for which all right-hand sides of three RGEs vanish simultaneously. As a result, there are three ?xed points and one ?xed line in the physical region and one 0xed point in the unphysical region (i.e., a ¡ 0 and X ¡ 0). Those in the physical region are given by [107] (X1∗ ; a∗1 ; G1∗ ) = (0; any; 0) ; (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0) ; 3 1 ∗ ∗ ∗ ; ;0 ; (X3 ; a3 ; G3 ) = 5 3 √ √ √ 87) 87) 2(2 + 45 2(11919 − 176 (X4∗ ; a∗4 ; G4∗ ) = ; 87; ; 4097 1069317 and it in the unphysical region is given by √ √ √ 87) 87) 2(2 − 45 2(11919 + 176 ; − 87; : (X5∗ ; a∗5 ; G5∗ ) = 4097 1069317

(4.260)

(4.261)

We should note that G = 0 is a 0xed point of the RGE for G, and a = 1 is the one for a. Hence RG >ows on G = 0 plane and a = 1 plane are con0ned in the respective planes. Now, let us study the phase structure of the HLS. Below we shall 0rst study the phase structure on G = 0 plane, second on a = 1 plane, and then on whole (X; a; G) space. Note also that RG >ows

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X 2 1.5 1 0.5

-0.5

0.5

1

1

1.5

2

2

2.5

3

a

Fig. 12. Phase diagram on G = 0 plane. Arrows on the >ows are written from the ultraviolet to the infrared. Gray line denotes the 0xed line (X1∗ ; a∗1 ; G1∗ ) = (0; any; 0). Points indicated by ⊕ and ⊗ (VM point; see Section 6) denote the 0xed points (3=5; 1=3; 0) and (1; 1; 0), respectively. Dashed lines divide the broken phase (lower side) and the symmetric phase (upper side; cross-hatched area): Flows drawn by thick lines are in the broken phase, while those by thin lines are in the symmetric phase.

in the region of X ¡ 0 (unphysical region) is con0ned in that region since X = 0 is the 0xed point of the RGE for X in Eq. (4.256). We 0rst study the phase structure of the HLS for G(m )=0 (g2 (m )=0). In this case m vanishes and the RGEs (4.256)–(4.258) are valid all the way down to the low energy limit, $ ¿ m = 0. Then the conditions in Eq. (4.251) are rewritten into the following conditions for X (0): (A-i) X (0) = 0 (m = 0) ⇔ broken phase ; (A-ii) X (0) = 0 (m = 0) ⇔ symmetric phase :

(4.262)

We show the phase diagram on G =0 plane in Fig. 12. There are one 0xed line and two 0xed points: (X1∗ ; a∗1 ; G1∗ ), (X2∗ ; a∗2 ; G2∗ ) and (X3∗ ; a∗3 ; G3∗ ). As we showed in Eq. (4.262), the phase is determined by the value of X ($) at the infrared limit $ = 0. In particular, the phase boundary is speci0ed by F2 (0) = 0, namely, governed by the infrared 0xed point such that X (0) = 0 [see Eq. (4.262)]. Such a 0xed point is the point (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0), which is nothing but the point corresponding to the vector manifestation (VM) [106] (see Section 6). Then the phase boundary is given by the RG >ows entering (X2∗ ; a∗2 ; G2∗ ). Since a = 1=3 is a 0xed point of the RGE for a in Eq. (4.257) for G = 0, the RG >ows for a ¡ 1=3 cannot enter (X2∗ ; a∗2 ; G2∗ ). Hence there is no phase boundary speci0ed by F2 (0) = 0 in a ¡ 1=3 region. Instead, F2 (0) vanishes even though F2 (0) = 0, namely a(0) = X (0) = 0. Then the phase boundary for a ¡ 1=3 is given by the RG >ow entering the point (X; a; G) = (0; 0; 0). In Fig. 12 the phase boundary is drawn by the dashed line, which divides the phases into the symmetric phase 37 (upper side; cross-hatched area) and the broken one (lower side). Here we should stress that the exact G ≡ 0 plane does not actually correspond to the underlying QCD as we shall demonstrate in Sections 6.1.4 and 6.1.5 and hence Fig. 12 is only for illustration of the section at G = 0 of the phase diagram in entire parameter space (X; a; G). 37

Here “symmetric phase” means that F2 ($) = 0 or F2 ($) = 0, namely 1=X ($) = F2 ($)=C$2 = 0 or a($) = 0 for non-zero (0nite) $.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

105

X

2 1.5 1 0.5

0.5

1

1.5

2

G

Fig. 13. Phase diagram on a = 1 plane. Arrows on the >ows are written from the ultraviolet to the infrared. Point indicated by ⊗ denotes the VM 0xed point (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0). (See Section 6.) Flows drawn by thick lines are in the broken phase, while those by thin lines are in the symmetric phase (cross-hatched area). Dot-dashed line corresponds to the on-shell condition G = X . In the shaded area the RGEs (4.256)–(4.258) are not valid since has already decoupled.

In the case of G(m ) ¿ 0 (g2 (m ) ¿ 0), on the other hand, the generally becomes massive (m = 0), and thus decouples at m scale. As we said in Section 4.8, below the m scale a and G = g2 · Nf =[2(4)2 ] no longer run, while F still runs by the loop e/ect. The running of F for $ ¡ m (denoted by F() ) is given in Eq. (4.219). From this we should note that the quadratic divergence (second term in Eq. (4.219)) of the loop can give rise to chiral symmetry restoration F() (0) = 0 [104,107]. The resultant relation between the order parameter F2 (0) and the F2 (m ) is given by Eq. (4.222), which in terms of X (m ) is rewritten as Nf F2 (0) = m2 [X −1 (m ) − 2 + a(m )] : (4.263) 2(4)2 Thus, the phase is determined by the following conditions: (B-i) X −1 (m ) ¿ 2 − a(m ) (m ¿ 0) ⇔ broken phase ; (B-ii) X −1 (m ) = 2 − a(m )(m ¿ 0) ⇔ symmetric phase : Then, the phase boundary is speci0ed by the condition 1 : 2 − a(m ) = X (m )

(4.264) (4.265)

Combination of this with the on-shell condition in Eq. (4.259) determines a line, which is nothing but an edge of the phase boundary surface: The phase boundary surface is given by the collection of the RG >ows entering points on the line speci0ed by Eqs. (4.259) and (4.265) [107]. We now study the a=1 plane (see Fig. 13). The >ows stop at the on-shell of (G=X ; dot-dashed line in Fig. 13) and should be switched over to RGE of F() ($) as mentioned above. From Eqs. (4.259) and (4.265) with a = 1 the >ow entering (X; G) = (1; 1) (dashed line) is the phase boundary which distinguishes the broken phase (lower side) from the symmetric one (upper side; cross-hatched area). For a ¡ 1, RG >ows approach to the 0xed point (X3∗ ; a∗3 ; G3∗ ) = (3=5; 1=3; 0) in the idealized high energy limit ($ → ∞).

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0.8

a(µ)

X(µ)

1

0.6 0.4

10 5

0.2 0 (a)

2

4

6

8

10

12

14

0

log (µ / m ρ)

2

4

6

8

10

12

14

log (µ / m ρ)

(b)

0.4

G(µ)

0.3 0.2 0.1

0 (c)

2

4

6

8

10

12

14

log (µ / m ρ)

Fig. 14. Scale dependences of (a) X ($), (b) a($) and (c) G($) in QCD with Nf = 3. Shaded area denotes the physical region, m 6 $ 6 . Flow shown by the dashed line are obtained by extending it to the (unphysical) infrared region by taking literally the RGEs in Eq. (4.256)–(4.258). In an idealized high energy limit the >ow approaches to the 0xed point (X4∗ ; a∗4 ; G4∗ ) (0:2; 9:3; 0:02).

For a ¿ 1, RG >ows in the broken phase approach to (X4∗ ; a∗4 ; G4∗ ) (0:2; 9:3; 0:02), which is precisely the 0xed point that the RG Dow of the Nf = 3 QCD belongs to. To see how the RG >ow of Nf = 3 QCD approaches to this 0xed point, we show the $-dependence of X ($), a($) and G($) in Fig. 14 where values of the parameters at $ = m are set to be (X (m ); a(m ); G(m )) (0:46; 1:22; 0:38) through Wilsonian matching with the underlying QCD [105] [see Section 5]. The values of X close to 1=2 in the physical region (m 6 $ 6 ) are very unstable against RGE Dow, and hence X ∼ 1=2 is realized in a very accidental way. We shall return to this point in Section 6.3.4. Finally, we show the phase boundary surface in the whole (X; a; G) space in Fig. 15 from three di/erent view points. This shows that the phase boundary spreads in a wide region of the parameter space. When we take the HLS model literally, the chiral symmetry restoration can occur at any point on this phase boundary. However, when we match the HLS with the underlying QCD, only the point (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0), VM point, on the phase boundary surface is selected, since the axialvector and vector current correlators in HLS can be matched with those in QCD only at that point [106] (see Section 6). Here again we mention that as we will discuss in Section 6.1.5, we should consider the VM only as a limit (“VM limit”) with the bare parameters approaching the VM 0xed point from the broken

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a

3

2

1

0 8

6

8 6

4

4 0

X

2 1

a (a)

2

0 6 2

4 30

2

4 2

G (b)

a 3 8

2

1

X

0

6

0

G

0

6

X

4

2

0 6

(c)

4

G

2

0

Fig. 15. Phase boundary surface from three di/erent view points. Points indicated by ⊕ and ⊗ (VM point) denote the 0xed points (3=5; 1=3; 0) and (1; 1; 0), respectively. Gray line denotes the line speci0ed by Eqs. (4.259) and (4.265).

phase: (X (); a(); G()) → (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0), particularly G() → 0. Setting G() ≡ 0 would contradict the symmetry of the underlying QCD (see Section 6.1.4.) Also note that, since the VM 0xed point is not an infrared stable 0xed point as can be seen in Figs. 12 and 13, the parameters in the infrared region do not generally approach this 0xed point: In the case of G = 0 with (X (); a()) → (1; 1), we can easily see from Fig. 12 that the infrared parameters behave as (X (0); a(0)) → (0; 1). In the case of a = 1 with (X (); G()) → (1; 0), on the other hand, without extra 0ne tuning, we expect that G() leads to G(m ). This together with the on-shell condition in Eq. (4.259) implies that X (m ) =

m2 →0 ; F2 (m2 )

which will be explicitly shown by solving the RGEs later in Section 6.3.2.

(4.266)

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(Q 2 ) − Q 4 d ΠA 2 dQ

Q 2 = µ2

ChPT

QCD

HLS

F π2(0) m ρ2

Λ2

µ2

Fig. 16. Schematic view of our matching procedure: In the region “QCD” the solid line shows −Q4 (dTA (Q2 )=dQ2 )|Q2 =$2 calculated from the OPE in QCD, while in the region “HLS” it shows F2 ($) determined by the RGE (4.208), and in the region “ChPT” it shows [F() ($)]2 in Eq. (4.219). [See text for details.]

5. Wilsonian matching In the previous section we derived the renormalization group equations (RGEs) in the Wilsonian sense for several parameters of the HLS. In the RGEs we included the quadratic divergence in addition to the logarithmic divergences. In Ref. [105] it was shown that quadratic divergences have the physical meaning of phenomenological relevance besides phase transition, when we match the bare theory of the HLS with the underlying QCD (“Wilsonian matching”). In this section we review the Wilsonian matching proposed in Ref. [105]. Let us explain our basic strategy of the Wilsonian matching using the axialvector current correlator TA (Q2 ) de0ned in Eq. (4.223) or (5.1). We plot the $-dependence of −Q4 (dTA (Q2 )=dQ2 )|Q2 = $2 in Fig. 16. In the high energy region this current correlator can be calculated from the operator product expansion (OPE) in QCD. As we shall show in Section 5.1 the $-dependence (for Nc = 3) is determined by the main term 1 + #s = as ($2 =82 )(1 + #s ($)=), which is plotted in the region indicated by “QCD” in Fig. 16. At the scale around 1 GeV, which we call the matching scale, we integrate out the quarks and gluons since they are not well-de0ned degrees of freedom in the low energy region. We assume that by integrating out the quarks and gluons we obtain the bare Lagrangian of the e/ective 0eld theory, i.e., the HLS. This Lagrangian includes the hadrons lighter than the matching scale which are well-de0ned degrees of freedom in the low energy region. Note that, as we discussed in Sections 4.1 and 4.5, for the consistency of the systematic derivative expansion in the HLS the matching scale must be smaller than the chiral scale , =4F ()= Nf determined from the bare parameter F (). When the momentum is around the matching scale, Q2 ∼ 2 , the current correlator is well described by the tree contributions with including O(p4 ) terms. Then we have −Q4 (dTA(HLS) (Q2 )=dQ2 )|Q2 =2 = F2 (), where F2 () is the bare parameter of the Lagrangian corresponding to the decay constant. The current correlator below is calculated from the bare HLS Lagrangian de0ned at by including loop corrections with the e/ect of quadratic divergences. Then, we expect that

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−Q4 (dTA(HLS) (Q2 )=dQ2 )|Q2 =$2 is dominated by F2 ($). The running of F2 ($) is determined by the RGE (4.208), and it is shown by the line in the region indicated by “HLS” in Fig. 16. The important point here is that the bare parameter F2 () is determined by matching it with the current correlator in OPE, as we shall show in Section 5.1. In the present procedure we equate −Q4 (dTA(QCD) (Q2 )=dQ2 )|Q2 =2 with F2 (), so that the line in the region “QCD” connects with the line in the region “HLS”. At the scale of mass m , decouples. Then, we expect that −Q4 (dTA (Q2 )=dQ2 )|Q2 =$2 is dominated by [F() ($)]2 which runs by the e/ect of quadratic divergence from the -loop e/ect alone as shown in Eq. (4.219). The solid line in the region indicated by “ChPT” shows the $-dependence of [F() ($)]2 . Since the ordinary ChPT without HLS is not applicable around the mass scale m , the solid line in the “ChPT” does not connect with the one in the “HLS”. The di/erence is understood as the e/ect of the 0nite renormalization at the scale $ = m as shown in Eq. (4.221) or Eq. (4.236). Through the procedure, which we called the “Wilsonian matching” in Ref. [105], the physical quantity F2 (0) is related to the current correlator calculated in the OPE in QCD. In Section 5.1, we introduce the Wilsonian matching conditions which are derived by matching the vector and axialvector correlators in the HLS with those obtained by the OPE in QCD. Then, we determine the bare parameters of the HLS using the Wilsonian matching conditions in Section 5.2. Physical predictions are made in Section 5.3 for Nf = 3 QCD which is close to the real world. In Section 5.4, we consider QCD with Nf = 2 to show how the Nf -dependences of the physical quantities appear. Finally, in Section 5.5, we study the sum rules related to the vector and axialvector current correlators. 5.1. Matching HLS with the underlying QCD As is well known, the parameters in the bare theory can be identi0ed with that at the cuto/ scale in the Wilsonian renormalization scheme. In this subsection following Ref. [105] we will present a way to determine the bare parameters of the HLS by matching the axialvector and vector current correlators in the HLS with those obtained by the operator product expansion (OPE) in QCD. This is contrasted with the usual renormalization where the bare theory is never referred to. Let us start with the two-point functions of the non-singlet axialvector and vector currents:

a b i d 4 x eipx 0|TJ5$ (x)J5 (0)|0 = ab (p$ p − g$ p2 )TA (Q2 ) ;

i

d 4 x eipx 0|TJ$a (x)Jb (0)|0 = ab (p$ p − g$ p2 )TV (Q2 ) ;

(5.1)

where Q2 = −p2 . In the HLS these two-point functions are well described by the tree contributions with including O(p4 ) terms when the momentum is around the matching scale, Q2 ∼ 2 . By combining O(p4 ) terms in Eq. (4.27) with the leading terms in Eq. (4.20) the correlators in the HLS are given by [105] F 2 () TA(HLS) (Q2 ) = 2 − 2z2 () ; (5.2) Q F 2 () TV(HLS) (Q2 ) = 2 [1 − 2g2 ()z3 ()] − 2z1 () ; (5.3) M () + Q2

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where we de0ned M2 () ≡ g2 ()F2 () : The same correlators are evaluated by the OPE up until O(1=Q6 ) [171,172]: 3(Nc2 − 1) #s Q2 Nc 1 (QCD) 2 − 1+ ln 2 (Q ) = 2 TA 8 3 8Nc $

1 #s qq 2 (#s =)G$ G $ 3 96(Nc2 − 1) 1 S 2 + ; + + Nc Q4 Nc Nc2 2 3Nc Q6 TV(QCD) (Q2 ) =

3(Nc2 − 1) #s Q2 − 1+ ln 2 8Nc $

1 #s qq 2 (#s =)G$ G $ 3 96(Nc2 − 1) 1 S 2 − ; + − Nc Q4 Nc Nc2 2 3Nc Q6 1 82

(5.4)

(5.5)

Nc 3

(5.6)

where $ is the renormalization scale of QCD and we wrote the Nc -dependences explicitly (see, e.g., Ref. [28]). We require that current correlators in the HLS in Eqs. (5.2) and (5.3) can be matched with those in QCD in Eqs. (5.5) and (5.6). Of course, this matching cannot be made for any value of Q2 , since the Q2 -dependences of the current correlators in the HLS are completely di/erent from those in the OPE: In the HLS the derivative expansion (in positive power of Q) is used, and the expressions for the current correlators are valid in the low energy region. The OPE, on the other hand, is an asymptotic expansion (in negative power of Q), and it is valid in the high energy region. Since we calculate the current correlators in the HLS including the 0rst non-leading order [O(p4 )], we expect that we can match the correlators with those in the OPE up until the 0rst derivative. Note that both TA(QCD) and TV(QCD) explicitly depend on $. 38 Such dependences are assigned to the parameters z2 () and z1 (). This situation is similar to that for the parameters Hi in the ChPT [79,80] [see, e.g., Eq. (2.43)]. However, the di/erence between two correlators has no explicit dependence on $. Thus our 0rst Wilsonian matching condition is given by [105] F2 () F2 () [1 − 2g2 ()z3 ()] − 2[z2 () − z1 ()] − 2 2 + M2 () =

38

4(Nc2 − 1) #s qq S 2 : Nc2 6

(5.7)

It should be noticed that the #s = term and #s qq S 2 term in the right-hand sides of the matching conditions [Eqs. (5.7)–(5.9) depend on the renormalization point $ of QCD, and that those generate a small dependence of the bare parameters of the HLS on $. This $ is taken to be the matching scale in the QCD sum rule shown in Refs. [171,172]. Here we take $ to be equal to the matching scale .

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We also require that the 0rst derivative of TA(HLS) in Eq. (5.2) matches that of TA(QCD) in Eq. (5.5), and similarly for TV ’s in Eqs. (5.3) and (5.6). This requirement gives the following two Wilsonian matching conditions [105]: 39 Nc F2 () 1 (QCD) 2 d 2 (1 + A ) ; = −Q TA (Q ) = 2 2 2 dQ 8 3 Q2 =2 A ≡

3(Nc2 − 1) #s 22 (#s =)G$ G $ + 8Nc Nc 4 S 2 #s qq 288(Nc2 − 1) 1 1 + ; + Nc3 2 3Nc 6

(5.8)

F2 () 4 [1 − 2g2 ()z3 ()] (QCD) 2 d 2 = −Q TV (Q ) 2 2 2 2 2 dQ [ + M ()] Q2 =2 Nc 1 (1 + V ) ; = 2 8 3 V ≡

3(Nc2 − 1) #s 22 (#s =)G$ G $ + 8Nc Nc 4 1 #s qq S 2 288(Nc2 − 1) 1 − − : Nc3 2 3Nc 6

(5.9)

The above three equations (5.7)–(5.9) are the Wilsonian matching conditions proposed in Ref. [105]. These determine several bare parameters of the HLS without much ambiguity. Especially, the second condition (5.8) determines the ratio F ()= directly from QCD. It should be noticed that the above Wilsonian matching conditions determine the absolute value and the explicit dependence of bare parameters of HLS on the parameters of underlying QCD such as Nc (not just scaling properties in the large Nc limit) and QCD , which would never have been obtained without matching and in fact has never been achieved for the EFT before. Now we discuss the large Nc behavior of the bare parameters: As we will show explicitly in Section 6.3, it is natural to assume that the matching scale has no large Nc -dependence. Then, condition (5.8), together with the fact that each term in A in Eq. (5.8) has only small Nc -dependence, 40 shows that the bare parameter F2 () scales as Nc . This is consistent with the ordinary large Nc counting of the on-shell decay constant, F2 (0) ∼ Nc . In the Wilsonian matching condition (5.9) it is plausible to assume that the bare mass parameter M () does not scale in the large Nc since the on-shell mass m does not. The second term inside the square bracket in the numerator of the left-hand side, g2 ()z3 (), cannot increase with increasing Nc for the consistency with the chiral counting, and then we require that this does not 39

One might think that there appear corrections from and/or loops in the left-hand sides of Eqs. (5.8) and (5.9). However, such corrections are of higher order in the present counting scheme, and thus we neglect them here. 40 Note that #s scales as 1=Nc in the large Nc counting, and that both (#=)G$ G $ and qq S scale as Nc .

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have the large Nc scaling. These scaling properties together with the fact that the right-hand side of Eq. (5.9) scales as Nc imply that the bare parameter F2 () scales as Nc , and then the bare parameter a() = F2 ()=F2 () does not have large Nc dependence. Noting that M2 () = a()g2 ()F2 (), we see, from the scaling properties of F2 (), a() and √ M2 () determined above, that the HLS gauge coupling g() scales as 1= Nc which is consistent with the fact that g is the coupling of the interaction among three mesons. This scaling property of g() with the requirement that g2 ()z3 () does not have large Nc dependence leads to z3 () ∼ O(Nc ). Finally, in the Wilsonian matching condition (5.7) the 0rst and second terms in the left-hand side as well as the right-hand side scale as Nc , so that z2 () − z1 () also scales as Nc . To summarize the Wilsonian matching conditions lead to the following large Nc scaling properties of the bare parameters of the HLS Lagrangian: F () ∼ O( Nc ) ; a() ∼ O(1) ; g() ∼ O(1= Nc ) ; z3 () ∼ O(Nc ) ; z2 () − z1 () ∼ O(Nc ) :

(5.10)

Note that the above scaling properties under the large Nc can be also obtained by counting the number of traces in the Lagrangian as was done in Ref. [80] to determine the scaling properties of the low-energy constants of the ordinary chiral perturbation theory. 5.2. Determination of the bare parameters of the HLS Lagrangian In this subsection we determine the bare parameters related to the two-point functions of the axialvector and vector current correlators from QCD through the Wilsonian matching conditions shown in the previous subsection. The right-hand sides in Eqs. (5.7)–(5.9) are directly determined from QCD. First note that the matching scale must be smaller than the mass of a1 meson which is not included in our e/ective theory, whereas has to be big enough for the OPE to be valid. Here we use a typical value: = 1:1 GeV :

(5.11)

In order to check the sensitivity of our result to the input value we also study the cases for the following wide range of the values: = 1:0 ∼ 1:2 GeV :

(5.12)

For de0niteness of the proceeding analysis let us 0rst determine the current correlators from the OPE. For the value of the gluonic condensate we use "# # s G$ G $ = 0:012 GeV4 (5.13)

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shown in Ref. [171,172] as a typical value. In Ref. [78] the value of quark condensate is estimated as 41 qq S 1 GeV = −(225 ± 25 MeV)3 :

(5.14)

We use the center value and study the dependence of the result on the quark condensate by including the error shown above. There are some ambiguities for the value of QCD (see, e.g., Ref. [47]). Here we use QCD = 400 MeV ;

(5.15)

but again we also study the cases QCD = 300; 350; 400; 450 MeV ;

(5.16)

to check the sensitivity of our result to the input value. Furthermore, we use the one-loop running to estimate #s () and qq S : 4 #s () = ; L0 ln(2 =2QCD ) #s (1 GeV) A=2 qq S = qq S 1 GeV ; (5.17) #s () where L0 = A=

11Nc − 2Nf ; 3 9(Nc2 − 1) 3C2 : = L0 Nc (11Nc − 2Nf )

(5.18)

Note that our typical choice = 1:1 GeV and QCD = 400 MeV corresponds to #s ( = 1:1 GeV; QCD = 400 MeV) 0:69 :

(5.19)

From the above inputs we evaluate the current correlators in the OPE, and have for Nc = Nf = 3: A=V = 0:220 + 0:054 + (0:089)=(−0:057) ∼ 0:363=0:217 for the respective terms #s =, 22 #s G$ G $ 3 4

and

1408 #s qq S 2 27 6 3

$

896 #s qq S 2 − 27 6 3

(5.20) ;

appearing in the right-hand sides of the Wilsonian matching conditions (5.8) and (5.9). It implies that the terms 1 and #s = (0rst term of A=V ) give dominant contributions over the gluonic and the quark condensate terms in the right-hand sides of Eqs. (5.8) and (5.9). We also list in Table 4 the results for other parameter choices of QCD and together with the ambiguities coming from that of the quark condensate shown in Eq. (5.14). While the gluonic 41

In the previous paper [105] we used the SVZ value [171,172] qq S 1; GeV = −(250 MeV)3 and hence the numerical analysis here is slightly di/erent from the previous one, although consistent with it within the error.

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Table 4 Values of the terms of the axialvector and vector current correlators derived from the OPE QCD

#s =

(GG)

(qq-A) S

(qq-V) S

0.30

1.00 1.10 1.20

0.185 0.171 0.160

0.079 0.054 0.038

0:122 ± 0:081 0:068 ± 0:045 0:040 ± 0:027

−0:077 ± 0:052 −0:043 ± 0:029 −0:026 ± 0:017

0.35

1.00 1.10 1.20

0.212 0.194 0.180

0.079 0.054 0.038

0:140 ± 0:093 0:078 ± 0:052 0:046 ± 0:031

−0:089 ± 0:059 −0:050 ± 0:033 −0:029 ± 0:020

0.40

1.00 1.10 1.20

0.243 0.220 0.202

0.079 0.054 0.038

0:160 ± 0:107 0:089 ± 0:060 0:053 ± 0:035

−0:102 ± 0:068 −0:057 ± 0:038 −0:033 ± 0:022

0.45

1.00 1.10 1.20

0.278 0.249 0.227

0.079 0.054 0.038

0:183 ± 0:122 0:102 ± 0:068 0:060 ± 0:040

−0:117 ± 0:078 −0:065 ± 0:043 −0:038 ± 0:026

$ 22 (#s =)G$ G , and those in the 0fth [indicated by 3 4 #s qq S 2 S 2 3 896 #s qq 3 1408 and − , respectively. Units of QCD 27 27 6 6 3

Values in the fourth column indicated by (GG) are the values of

(qq-A)] S and the sixth [indicated by (qq-V)] S columns are of and are GeV. Errors in 0fth and sixth columns are from the error in the quark condensate qq S = −(225 ± 25 MeV) shown in Eq. (5.14).

condensate gives very small correction for any choice of the matching scale, the quark condensate gives a non-negligible correction for small matching scale ( 1 GeV). Now that we have determined the current correlators in the OPE, we can determine the bare parameters of the HLS through the Wilsonian matching conditions. Especially, the Wilsonian matching condition (5.8) determines directly the value of the bare decay constant F (). Before discussing details, we here give a rough estimation to get an essential point of our analysis: 2 Nc 2 Nc 2 2(1 + A ) ∼ 3 ; (5.21) F () = 2 (4) 3 4 3 where A was estimated in Eq. (5.20) for Nc = Nf = 3 and very roughly A ∼ 0:5 :

(5.22)

Note again that each term in Eq. (5.20) for A is rather independent of Nc . First of all Eq. (5.21) implies the derivative expansion parameter can be very small in the large Nc limit (with 0xed Nf ): 2 Nf Nf 3 Nf = ∼ 1 (Nc 1) : (5.23) 4F () 3 Nc Nc As we discussed in the previous section, we make the systematic expansion in the large Nc limit, and extrapolate the results to the real world. In QCD with Nc = Nf = 3 the above expansion parameter becomes of order one, so that one might think that the systematic expansion breaks down. However, as can be seen in, e.g., Eq. (4.181) with a ∼ 1, the quadratically divergent loop contributions to F2

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get an extra factor 1=2 due to the additional loop and hence the loop expansion would be valid up till 4F () ∼ ; (5.24) Nf =2 or Nf ∼1 : (5.25) 2Nc Furthermore, as we will show below in this section, the analysis based on the systematic expansion reproduces the experiment in good agreement. This shows that the extrapolation of the systematic expansion from the large Nc limit to the real world works very well. Now, by choosing the matching scale as = 1:1 GeV, or =4 86:4 MeV, the value of F () is estimated as F2 () ∼ 3(86:4 MeV)2 ∼ (150 MeV)2 : F2 ()

(5.26)

Then the Wilsonian matching predicts in terms of the QCD parameters and the value de0nitely disagrees with the on-shell value 86:4 ± 9:7 MeV in the chiral limit [79,81]. Were it not for the quadratic divergence, we would have met with a serious discrepancy between the QCD prediction and the physical value! How does the quadratic divergence save the situation? The key is the Wilsonian RGE derived in Section 4.8 which incorporated quadratic divergence (as well as logarithmic one) for the running of F2 . To perform a crude estimate let us neglect the e/ect of logarithmic divergence (by taking g() → 0) and include the e/ect of quadratic divergence only in the RGE for F2 in Eq. (4.208). Furthermore, as it turns out that the Wilsonian matching implies the bare value a() 1, we take a = 1, which is the 0xed point of RGE, so that the analytical solution of RGE becomes very simple: F2 (m ) = F2 () − (Nf =2(4)2 )(2 − m2 ). This together with the relation (4.222) yields the approximate relation between the bare parameter F2 () and the on-shell decay constant F2 (0) as

Nf Nf 2 Nc 2 2 2 (1 + A ) − ∼ 2 F (0) ∼ F () − 2(4)2 8 3 4 2 1 ∼ F2 () ∼ (100 MeV)2 ; (5.27) ∼ 1:5 4 2 where we adopted Nc = Nf = 3 and A ∼ 0:5 to obtain the last line. Then, the on-shell decay constant F (0) is now close to the value F (0)=86:4±9:7 MeV. The small deviation from 86:4 MeV will be resolved by taking account of the logarithmic correction with g() = 0 and the correction by a() = 1 (and more precise value A ∼ 0:363) for the realistic case Nc = Nf = 3. At any rate this already shows that the Wilsonian matching works well and quadratic divergence plays a vital role. Let us now determine the precise value of F () for given values of QCD and the matching scale in the case of Nc = Nf = 3. We list the resultant values of F () obtained from the Wilsonian matching condition (5.8) together with the ambiguity from that of the quark condensate qq S = −(225 ± 25 MeV)3 in Table 5. This shows that the bare decay constant is determined from the matching condition without much ambiguity: It is almost determined by the main term 1+#s = in the right-hand side of Eq. (5.8), and the ambiguity of the quark condensate qq =−(225±25 S MeV)3 shown in Eq. (5.14) does not a/ect to the bare decay constant very much.

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Table 5 Values of the bare decay constant F () determined through the Wilsonian matching condition (5.8) for given QCD and the matching scale QCD

F ()

QCD

F ()

0.30

1.00 1.10 1:20

0:132 ± 0:004 0:141 ± 0:002 0:150 ± 0:002

0.40

1.00 1.10 1.20

0:137 ± 0:005 0:145 ± 0:003 0:154 ± 0:002

0.35

1.00 1.10 1.20

0:135 ± 0:004 0:143 ± 0:003 0:152 ± 0:002

0.45

1.00 1.10 1.20

0:140 ± 0:006 0:147 ± 0:004 0:155 ± 0:002

Units of QCD , and F () are GeV. Note that error of F () is from the error in the quark condensate qq S = −(225 ± 25 MeV)3 shown in Eq. (5.14).

There are four parameters a(), g(), z3 () and z2 () − z1 () other than F (), which are relevant to the low energy phenomena related to two correlators analyzed in the previous subsection. 42 We have already used one Wilsonian matching condition (5.8) to determine one of the bare parameters F () for a given matching scale . The remaining two Wilsonian matching conditions in Eqs. (5.7) and (5.9) are not enough to determine other four relevant bare parameters. We therefore use the on-shell pion decay constant F (0) = 86:4 ± 9:7 MeV estimated in the chiral limit [79–81] and the mass m = 771:1 MeV as inputs: We chose a() and g() which, combined with F () determined from the Wilsonian matching condition (5.8), reproduce F (0) and m through the Wilsonian RGEs in Eqs. (4.208), (4.210) and (4.211). Then, we use the matching condition (5.9) to determine z3 (). Finally z2 () − z1 () is 0xed by the matching condition (5.7). The resultant values of 0ve bare parameters of the HLS are shown in Tables 6 and 7 for = 1:0, 1:1 and 1:2 GeV. Typical values of the bare parameters for (QCD ; ) = (0:40; 1:10) GeV are F () = 145 ± 3 MeV ; a() = 1:33 ± 0:28 ± 0:14 ; g() = 3:69 ± 0:13 ± 0:02 ; z3 () = (−5:84 ± 5:78 ± 0:18) × 10−3 ; z2 () − z1 () = (−1:79 ± 0:34 ± 0:61) × 10−3 ;

(5.28)

where the error of F () comes only from qq S = −(225 ± 25 MeV)3 , while the 0rst error for a(), g(), z3 () and z2 () − z1 () comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = 3 2 2 −(225±25 MeV) . By using the above values, the bare mass de0ned by M ()=a()g ()F2 () 42

As we noted in the previous subsection, although each of z1 () and z2 () depends on the renormalization point $ of QCD, the di/erence z2 () − z1 () does not. Actually, z2 () + z1 () corresponds to the parameter Hi in the ChPT [79,80] [see H1 of Eq. (4.38)]. Thus, the di/erence z2 () − z1 () is relevant to the low energy phenomena, while z2 () + z1 () is irrelevant.

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Table 6 Leading order parameters of the HLS at $ = for several values of QCD and QCD

F ()

a()

g()

0.30

1.00 1.10 1.20

0:132 ± 0:004 0:141 ± 0:002 0:150 ± 0:002

1:41 ± 0:29 ± 0:16 1:49 ± 0:30 ± 0:11 1:49 ± 0:30 ± 0:08

4:05 ± 0:16 ± 0:01 3:68 ± 0:11 ± 0:00 3:42 ± 0:09 ± 0:00

0.35

1.00 1.10 1.20

0:135 ± 0:004 0:143 ± 0:003 0:152 ± 0:002

1:32 ± 0:28 ± 0:18 1:41 ± 0:29 ± 0:12 1:42 ± 0:29 ± 0:09

4:06 ± 0:18 ± 0:03 3:68 ± 0:12 ± 0:01 3:42 ± 0:10 ± 0:00

0.40

1.00 1.10 1.20

0:137 ± 0:005 0:145 ± 0:003 0:154 ± 0:002

1:22 ± 0:28 ± 0:21 1:33 ± 0:28 ± 0:14 1:34 ± 0:28 ± 0:09

4:09 ± 0:20 ± 0:06 3:69 ± 0:13 ± 0:02 3:43 ± 0:10 ± 0:01

0.45

1.00 1.10 1.20

0:140 ± 0:006 0:147 ± 0:004 0:155 ± 0:002

1:10 ± 0:28 ± 0:24 1:23 ± 0:27 ± 0:16 1:26 ± 0:26 ± 0:10

4:13 ± 0:22 ± 0:10 3:71 ± 0:14 ± 0:03 3:44 ± 0:11 ± 0:01

Units of QCD , and F () are GeV. The error of F () comes only from qq S = −(225 ± 25 MeV)3 . The 0rst error for a() and g() comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = −(225 ± 25 MeV)3 . Note that 0.00 in the error of g() implies that the error is smaller than 0.01.

Table 7 Two of next-leading order parameters of the HLS at $ = for several values of QCD and QCD

z3 ()

z2 () − z1 ()

0.30

1.00 1.10 1.20

−6:10 ± 4:36 ± 0:63 −3:14 ± 5:04 ± 0:19 −1:27 ± 5:92 ± 0:12

−2:21 ± 0:37 ± 0:84 −2:01 ± 0:34 ± 0:47 −1:76 ± 0:30 ± 0:27

0.35

1.00 1.10 1.20

−7:20 ± 4:73 ± 0:41 −4:35 ± 5:38 ± 0:04 −2:66 ± 6:22 ± 0:23

−2:05 ± 0:38 ± 0:97 −1:90 ± 0:34 ± 0:54 −1:69 ± 0:29 ± 0:31

0.40

1.00 1.10 1.20

−8:65 ± 5:19 ± 0:05 −5:84 ± 5:78 ± 0:18 −4:31 ± 6:56 ± 0:39

−1:85 ± 0:39 ± 1:12 −1:79 ± 0:34 ± 0:61 −1:61 ± 0:29 ± 0:35

0.45

1.00 1.10 1.20

−10:6 ± 5:79 ± 0:56 −7:73 ± 6:27 ± 0:52 −6:29 ± 6:96 ± 0:61

−1:61 ± 0:41 ± 1:29 −1:65 ± 0:35 ± 0:70 −1:52 ± 0:29 ± 0:40

Units of QCD and are GeV. Values of z3 () and z2 () − z1 () are scaled by a factor of 103 . The 0rst error comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = −(225 ± 25 MeV)3 .

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is estimated as M () = 614 ± 44 ± 16 MeV :

(5.29)

These values show that the ambiguities of the bare parameters coming from that of the quark condensate are small for the leading order parameters as well as the parameter z3 , while it is rather large for z2 − z1 . This is because the leading order parameters are almost determined by the 1 + #s = term of the current correlators derived from the OPE through the Wilsonian matching, while z2 − z1 is directly related to the quark condensate as in Eq. (5.7). Now, one might suspect that the inclusion of the A1 (a1 meson and its >avor partners) would a/ect the above matching result, since the mass of a1 is ma1 = 1:23 ± 0:04 GeV [91] close to our matching scale = 1:1 GeV. Especially, it might give a large contribution in determining the value of F2 () so as to pull it down close to the F2 (0) (86:4 MeV)2 , and hence the large amount of the quadratic divergence might be an artifact of simply neglecting the A1 contribution. However, this is not the case. Inclusion of A1 does not a/ect the large value of F (). This is seen as follows: We can include the e/ect of A1 by using the e/ective 0eld theory such as the Generalized HLS [17,23]. Although a complete list of the O(p4 ) terms has not yet been given, on the analogy of the contribution to the vector current correlator given in Eq. (5.3) it is reasonable to write the axialvector current correlator around the matching scale with the A1 contribution included as TA(GHLS) (Q2 ) =

FA21 () F2 () + − 2z2 () ; Q2 MA21 () + Q2

(5.30)

where MA1 () is the bare A1 mass, FA1 () the bare A1 decay constant analog to F () ≡ F2 ()[1 − 2g2 ()z3 ()] and z2 () exhibits the contribution from higher modes analog to z2 (). By using the above correlator, the Wilsonian matching condition (5.8) would be changed to 2 FA21 () Nc 1 F2 () (1 + A ) : + = 2 (5.31) 2 2 2 2 8 3 [MA1 () + ] For determining the value of F () from the above Wilsonian matching condition we need to know the values of MA1 () and FA1 (). Since the matching scale is close to the A1 mass, the on-shell values give a good approximation, MA1 () mA1 , FA1 () FA1 (mA1 ) = FA1 . Although the experimental value of the a1 mass is known as ma1 = 1:23 ± 0:04 GeV [91], the on-shell value of its decay constant Fa1 is not known. However, we could use the pole saturated version of the 0rst Weinberg’s sum rule [184] F2 − FA21 = F2 (0) ;

(5.32)

together with F = g =m = 0:154 ± 0:001 GeV and F (0) = 86:4 ± 9:7 MeV, which yields FA1 = 0:127 ± 0:007 GeV, and hence roughly FA21 () (130 MeV)2 . Here, instead of this value, we use FA21 () = F2 ∼ (150 MeV)2 ∼ 3(=4)2 and set MA1 () ∼ to include a possible maximal A1 contribution to the Wilsonian matching condition (5.31). The resultant value (a possible minimum value) of F () with Nc = 3 is estimated as F2 () =

4 FA21 () 2 (1 + ) − A 82 [MA21 () + 2 ]2

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

119

Table 8 Five parameters of the HLS at $ = m for several values of QCD with = 1:1 GeV. Units of QCD , and F are GeV QCD

F (m )

a(m )

g(m )

0.30 0.35 0.40 0.45

1.10 1.10 1.10 1.10

0:0995 ± 0:0012 ± 0:0036 0:102 ± 0:001 ± 0:004 0:105 ± 0:001 ± 0:004 0:108 ± 0:001 ± 0:005

1:57 ± 0:34 ± 0:13 1:48 ± 0:33 ± 0:14 1:38 ± 0:32 ± 0:16 1:27 ± 0:32 ± 0:18

6:19 ± 0:59 ± 0:03 6:22 ± 0:64 ± 0:06 6:27 ± 0:69 ± 0:11 6:36 ± 0:76 ± 0:17

QCD

z3 (m )

z2 (m ) − z1 (m )

0.30 0.35 0.40 0.45

1.10 1.10 1.10 1.10

−4:13 ± 5:20 ± 0:13 −5:38 ± 5:52 ± 0:01 −6:90 ± 5:89 ± 0:23 −8:83 ± 6:34 ± 0:56

−2:49 ± 0:59 ± 0:56 −2:32 ± 0:60 ± 0:64 −2:13 ± 0:61 ± 0:74 −1:89 ± 0:62 ± 0:86

Values of z3 (m ) and z2 (m ) − z1 (m ) are scaled by a factor of 103 . Note that the 0rst error comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = −(225 ± 25 MeV)3 .

2 3 2(1 + A ) − ∼ (4)2 4 2 3 9 2 × 86:4 MeV ∼ (130 MeV)2 ; ∼ ∼ 4 (4)2 2

(5.33)

where we again adopted a very rough estimate A ∼ 0:5 to obtain the last line. This value F () ∼ 130 MeV (possible minimum value) is still much larger than the on-shell value 86:4 ± 9:7 MeV and close to the value 150 MeV obtained in Eq. (5.26) by the Wilsonian matching without including the e/ect of A1 . 5.3. Results of the Wilsonian matching 5.3.1. Full analysis In the previous subsection we have completely speci0ed the bare Lagrangian through the Wilsonian matching conditions (5.7)–(5.9) together with the physical inputs of the pion decay constant F (0) and the rho mass m . Using the Wilsonian RGEs for the parameters obtained in Section 4.8 [Eqs. (4.208), (4.210), (4.211), (4.213)–(4.215)], we obtain the values of 0ve parameters at $ = m . In Table 8 we list several typical values of 0ve parameters at $ = m for several values of QCD with = 1:1 GeV. Typical values for (QCD ; ) = (0:40; 1:10) GeV are F (m ) = 105 ± 1 ± 4 MeV ; a(m ) = 1:38 ± 0:32 ± 0:16 ; g(m ) = 6:27 ± 0:69 ± 0:11 ;

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Table 9 Physical quantities predicted by the Wilsonian matching conditions and the Wilsonian RGEs QCD

g

g

a(0)

0.30

1.00 1.10 1.20

0:121 ± 0:009 ± 0:003 0:111 ± 0:011 ± 0:001 0:105 ± 0:014 ± 0:000

6:41 ± 0:78 ± 0:08 6:44 ± 0:83 ± 0:03 6:43 ± 0:82 ± 0:02

2:07 ± 0:04 ± 0:05 2:08 ± 0:07 ± 0:02 2:08 ± 0:06 ± 0:02

0.35

1.00 1.10 1.20

0:125 ± 0:011 ± 0:003 0:116 ± 0:013 ± 0:001 0:110 ± 0:015 ± 0:000

6:36 ± 0:72 ± 0:15 6:41 ± 0:78 ± 0:06 6:40 ± 0:78 ± 0:04

2:03 ± 0:00 ± 0:09 2:06 ± 0:04 ± 0:04 2:06 ± 0:04 ± 0:03

0.40

1.00 1.10 1.20

0:129 ± 0:013 ± 0:002 0:121 ± 0:014 ± 0:000 0:116 ± 0:017 ± 0:001

6:26 ± 0:63 ± 0:24 6:35 ± 0:72 ± 0:11 6:35 ± 0:74 ± 0:07

1:97 ± 0:04 ± 0:15 2:03 ± 0:01 ± 0:07 2:03 ± 0:02 ± 0:04

0.45

1.00 1.10 1.20

0:136 ± 0:016 ± 0:001 0:127 ± 0:017 ± 0:000 0:123 ± 0:019 ± 0:001 0:119 ± 0:001

6:10 ± 0:52 ± 0:38 6:26 ± 0:65 ± 0:17 6:28 ± 0:69 ± 0:10 6:00 ± 0:01

1:87 ± 0:10 ± 0:23 1:97 ± 0:03 ± 0:11 1:98 ± 0:01 ± 0:06

Exp.

Units of QCD and are GeV, and that of g is GeV2 . Experimental values of g and g are derived in Section 3.5. Note that the 0rst error comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = −(225 ± 25 MeV)3 . 0.000 in the error of g and 0.00 in a(0) imply that the errors are smaller than 0.001 and 0.01, respectively.

z3 (m ) = (−6:90 ± 5:89 ± 0:23) × 10−3 ; z2 (m ) − z1 (m ) = (−2:13 ± 0:61 ± 0:74) × 10−3 ;

(5.34)

where the 0rst error comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = −(225 ± 25 MeV)3 . It should be noticed that, comparing the above value of a(m ) with that of a() in Eq. (5.28), we see that the parameter a does not change its value by the running from the matching scale to the scale of on-shell. Furthermore, the value itself is close to one. Nevertheless, the parameter a at the low-energy limit becomes closer to 2 which leads to the vector dominance of the electromagnetic form factor of the pion [see the analysis around Eq. (5.45)]. Now that we have determined the 0ve parameters at $ =m , we make several physical predictions. The typical “physical” quantities derived from the 0ve parameters are [105] – mixing strength, Gasser–Leutwyler’s parameter L10 [80], –– coupling constant g , Gasser–Leutwyler’s parameter L9 [80] and the parameter a(0) which parameterizes the validity of the vector dominance. Below we shall list the relations of the 0ve parameters of the HLS to these “physical” quantities following Ref. [105]. The resultant predictions are listed in Tables 9 and 10 for several values of QCD and . – mixing strength. The second term in Eq. (4.20) gives the mass mixing between and the external 0eld of (photon 0eld). The z3 -term in Eq. (4.27) gives the kinetic mixing. Combining these two at the on-shell of leads to the – mixing strength [105]: g = g(m )F2 (m )[1 − g2 (m )z3 (m )] ;

(5.35)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

121

Table 10 Values of Gasser–Leutwyler’s parameters L9 and L10 predicted by the Wilsonian matching conditions and the Wilsonian RGEs QCD

L9 (m )

L10 (m )

0:30

1:00 1:10 1:20

6:88 ± 0:21 ± 0:30 6:24 ± 0:05 ± 0:09 5:82 ± 0:26 ± 0:01

−4:07 ± 1:94 ± 0:46 −2:62 ± 2:38 ± 0:43 −1:94 ± 2:75 ± 0:38

0:35

1:00 1:10 1:20

7:04 ± 0:22 ± 0:38 6:50 ± 0:05 ± 0:12 6:12 ± 0:27 ± 0:02

−4:87 ± 2:02 ± 0:56 −3:46 ± 2:45 ± 0:52 −2:84 ± 2:82 ± 0:44

0:40

1:00 1:10 1:20

7:22 ± 0:21 ± 0:48 6:77 ± 0:07 ± 0:16 6:45 ± 0:30 ± 0:03

−5:82 ± 2:13 ± 0:71 −4:43 ± 2:54 ± 0:62 −3:84 ± 2:91 ± 0:52

0:45

1:00 1:10 1:20

7:41 ± 0:17 ± 0:59 7:07 ± 0:10 ± 0:20 6:81 ± 0:35 ± 0:04

−6:98 ± 2:31 ± 0:93 −5:57 ± 2:68 ± 0:76 −4:99 ± 3:03 ± 0:63

6:5 ± 0:6

−5:1 ± 0:7

Exp.

Units of QCD and are GeV. Values of Lr9 (m ) and Lr10 (m ) are scaled by a factor of 103 . Experimental values of Lr9 (m ) and Lr10 (m ) are derived in Sections 2.8 and 2.9. Note that the 0rst error comes from F (0) = 86:4 ± 9:7 MeV and the second error from qq S = −(225 ± 25 MeV)3 .

which should be compared with the quantity derived from the experimental data of the → e+ e− decay width. As we have shown in Eq. (3.72), 5( → e+ e− ) = (6:85 ± 0:11) × 10−3 MeV [91] leads to g |exp = 0:119 ± 0:001 GeV2 . The typical predicted value of g for (QCD ; ) = (0:40; 1:10) GeV is g |theo = 0:121 ± 0:014 ± 0:0003 GeV2 ;

(5.36)

where the 0rst error comes from the ambiguity of the input value of F (0) and the second one from that of the quark condensate qq . S The central value of this as well as that for (QCD ; ) = (0:30; 1:00) GeV shown in Table 9 are very close to the experimental value. These values are improved from the tree prediction g |tree = 0:103 GeV2 in Eq. (3.76) where g in addition to F (0) and m was used as an input. It should be noticed that most predicted values are consistent with the experiment within the error of input values of F (0) and qq . S Gasser–Leutwyler’s parameter L10 [80]. As we have done in Section 4.9 the relation between the Gasser–Leutwyler’s parameter L10 and the parameters of HLS is obtained by matching the axialvector current correlator in the low energy limit. The resultant relation is given by [see Eq. (4.249)] 43 Nf 11a(m ) Nf 5 z3 (m ) − z2 (m ) + z1 (m ) 1 + + + : (5.37) Lr10 (m ) = − 2 2 (4) (4)2 72 4g (m ) 2 96 43

Note that the 0nite correction appearing as the last term in the right-hand side of Eq. (5.37) was not included in Ref. [105].

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

The ‘experimental’ value of L10 is estimated as [see Eq. (2.72) in Section 2.9] Lr10 (m )|exp = (−5:1 ± 0:7) × 10−3 . A typical value of the prediction is Lr10 (m )|theo = (−4:43 ± 2:54 ± 0:62) × 10−3 ;

(5.38)

for (QCD ; ) = (0:40; 1:10) GeV (see Table 10) where the 0rst error comes from the ambiguity of the input value of F (0) and the second one from that of the quark condensate. There are large ambiguities mainly from that of F (0), and all the predicted values shown in Table 10 are consistent with the experimental value. We should note that the central value of the prediction is somewhat improved from the tree value LV10 = (−7:4 ± 2:3) × 10−3 in Table 3 in Section 3.6. –– coupling constant g . Strictly speaking, we have to include a higher derivative type z4 -term listed in Eq. (4.27). However, a detailed analysis [101] using a similar model [128] does not require its existence. 44 Hence we neglect the z4 -term. If we simply read the –– interaction from Eq. (4.20), we would obtain g = g(m )F2 (m )=2F2 (m ). However, g should be de0ned for on-shell and ’s. While F2 and g2 do not run for $ ¡ m , F2 does run. The on-shell pion decay constant is given by F (0). Thus we have to use F (0) to de0ne the on-shell –– coupling constant. The resulting expression is given by [105] g =

g(m ) F2 (m ) : 2 F2 (0)

(5.39)

As we have shown in Eq. (3.71), the experimental value of g is estimated as g |exp =6:00±0:01. A typical value of the prediction is g |theo = 6:35 ± 0:72 ± 0:11 ;

(5.40)

for (QCD ; ) = (0:40; 1:10) GeV (see Table 9). The error of the prediction mainly comes from the ambiguity of the input of F (0), and all the predictions shown in Table 9 are consistent with the experiment. Note that, in the tree-level analysis done in Section 3.5, g was used as an input. Gasser–Leutwyler’s parameter L9 [80]. Similarly to the z4 -term contribution to g we neglect the contribution from the higher derivative type z6 -term. The resultant relation between L9 and the parameters of the HLS is given by [105,177] Nf 5 1 1 r = − z3 (m ) ; L9 (m ) + (5.41) (4)2 72 4 g2 (m ) where the second term in the left-hand side is the 0nite correction derived in the ChPT [79,80,177]. 45 The ‘experimental value’ of L9 is estimated as [see Eq. (2.66) in Section 2.8] Lr9 (m )|exp = (6:5 ± 0:6) × 10−3 , and the typical prediction is Lr9 (m )|theo = (6:77 ± 0:07 ± 0:16) × 10−3 ;

(5.42)

for (QCD ; ) = (0:4; 1:1) GeV. The ambiguity in the theoretical prediction from the input value of F (0) is not so large as that for L10 . But the experimental error is about 10%, so that most predictions are consistent with the experiment (see Table 10). 44

Note that the existence of the kinetic type – mixing from z3 -term was needed to explain the experimental data of 5( → e+ e− ) [101]. 45 This 0nite correction in the ChPT was not included in Ref. [105].

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

123

2

a ( µ)

1.5 1 0.5

0.2

0.4

0.6

µ (GeV)

0.8

1

Fig. 17. Running of a($) for (QCD ; )=(0:40; 1:10) GeV. Gap at m is due to the e/ect of 0nite renormalization between F2 (m ) and [F() (m )]2 given in Eq. (4.220).

Parameter a(0). We further de0ne the parameter a(0) by the direct –– interaction in the second term in Eq. (4.20). As we stated above, F2 does not run for $ ¡ m while F2 does. Thus we have [105,107] ($ ¿ m ) ; F2 ($)=F2 ($) a($) ≡ (5.43) 2 () 2 F (m )=[F ($)] ($ ¡ m ) : This parameter for on-shell pions becomes F 2 (m ) ; (5.44) a(0) = 2 F (0) which should be compared with the parameter a used in the tree-level analysis, a = 2 corresponding to the vector dominance (VD) [21,24]. Most values of the prediction are close to 2: We obtained a(0) 2 ;

(5.45)

although a() a(m ) 1. We show the running of a($) for (QCD ; ) = (0:40; 1:10) GeV in Fig. 17. This shows that although a($) 1 for m ¡ $ ¡ , a(0) 2 is realized by the running of [F() ($)]2 . KSRF relations. The KSRF (I) relation g = 2g F2 [126,163] holds as a low energy theorem of the HLS [22,23,95,96,103]. Here this is satis0ed as follows [105]: As we have shown in Section 4.7, higher derivative terms like z3 do not contribute in the low energy limit, and the – mixing strength becomes g (0) = g(m )F2 (m ). Comparing this with g in Eq. (5.39), 46 we can easily read that the low energy theorem is satis?ed. As to the on-shell , on the other hand, using the decay constant at the chiral limit, F (0) = 86:4 ± 9:7 MeV, together with the experimental values of the – mixing strength, g = 0:119 ± 0:001 GeV2 , and the –– coupling, g = 6:00 ± 0:01, we have g = 1:32 ± 0:30 : (5.46) 2g F2 (0) exp 46

The contribution from the higher derivative term is neglected in the expression of g given in Eq. (5.39), i.e., g = g (m2 ; 0; 0) = g (0; 0; 0).

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

This implies that there is about 30% deviation of the experimental value from the KSRF (I) relation. 47 As we have studied in Section 3.5, at the leading order this ratio is predicted as 1: (g =2g F2 (0))|tree = 1. When the next order correction generated by the loop e/ect and the O(p4 ) term is included, the combination of g in Eq. (5.35) with g in Eq. (5.39) provides g = 1 − g2 (m )z3 (m ) = 1:27 ± 0:29 ± 0:02 ; (5.47) 2g F2 (0) theo where the value is obtained for (QCD ; ) = (0:40; 1:1). This shows that the 30% deviation of experimental value from the KSRF (I) relation as in Eq. (5.46) is actually explained by the existence of the z3 term together with the loop e/ect included through the Wilsonian RGEs. 2 The KSRF (II) relation m2 =2g F2 [126,163] is approximately satis0ed by the on-shell quantities even though a(m ) 1. This is seen as follows [105]: Eq. (5.39) with Eq. (5.44) and m2 = 2 F2 (0) = m2 (a(0)=2). Thus a(0) 2 leads to the approximate KSRF g2 (m )F2 (m ) leads to 2g (II) relation. Furthermore, a(0) 2 implies that the direct -– coupling is suppressed (vector dominance). We shall return to this point later (see Section 6.3.4). To summarize, the predicted values of g , g , Lr9 (m ) and Lr10 (m ) remarkably agree with the experiment, although Lr10 (m ) is somewhat sensitive to the values of QCD and . 48 There are considerable ambiguities from the input value of F (0), and most predicted values are consistent with the experiment. Furthermore, we have a(0) 2, although a() a(m ) 1. The KSRF (I) relation is reproduced better than the tree-level result and KSRF (II) relation holds even for a(m ) 1. 5.3.2. “Phenomenology” with a() = 1 As we have seen above, a() = 1 is already close to the reality. Here it is worth emphasizing this fact by demonstrating more explicitly, since a() = 1 is a 0xed point of the RGE and of direct relevance to the Vector Manifestation we shall fully discuss in Section 6. We shall show the result of the same analysis as that already done above except a point that one of the input data, F (0) = 86:4 ± 9:7 MeV, is replaced by a() = 1. First, the bare parameters in the case a() = 1 for (; QCD ) = (1:1; 0:4) GeV are given by g() = 3:86 ± 0:04 ; z3 () = (−13:8 ± 2:8) × 10−3 ; z2 () − z1 () = (−1:37 ± 0:43) × 10−3 :

(5.48)

In Tables 11 and 12 we show the values of the bare parameters for several choices of and QCD . 47

Note that, in Eq. (3.73), we used the experimental value of the decay constant, F; phys = 92:42 ± 0:23, and obtained (g =2g F2 )|exp = 1:15 ± 0:01. Strictly speaking, we may have to include the e/ect of explicit chiral symmetry breaking due to the current quark masses into g as well as g . However, according to the analysis done by the similar model at tree level in Ref. [101], the corrections from the explicit chiral symmetry breaking to them are small. So we neglect the e/ect in the present analysis. 48 One might think of the matching by the Borel transformation of the correlators. However, agreement of the predicted values, especially g , are not as remarkably good as that for the present case.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

125

Table 11 The parameters F (), g() and M () in the case of a() = 1 for several values of QCD and QCD

F ()

g()

M ()

0:30

1:00 1:10 1:20

0:132 ± 0:004 0:141 ± 0:002 0:150 ± 0:002

4:33 ± 0:07 3:91 ± 0:03 3:61 ± 0:02

0:574 ± 0:008 0:551 ± 0:005 0:542 ± 0:003

0:35

1:00 1:10 1:20

0:135 ± 0:004 0:143 ± 0:003 0:152 ± 0:002

4:30 ± 0:07 3:89 ± 0:04 3:59 ± 0:02

0:578 ± 0:009 0:554 ± 0:006 0:545 ± 0:004

0:40

1:00 1:10 1:20

0:137 ± 0:005 0:145 ± 0:003 0:154 ± 0:002

4:26 ± 0:08 3:86 ± 0:04 3:57 ± 0:02

0:583 ± 0:010 0:558 ± 0:006 0:548 ± 0:004

0:45

1:00 1:10 1:20

0:140 ± 0:006 0:147 ± 0:004 0:155 ± 0:002

4:21 ± 0:09 3:84 ± 0:05 3:55 ± 0:02

0:588 ± 0:011 0:563 ± 0:007 0:552 ± 0:004

Units of QCD , , F () and M () are GeV. The errors come from qq S = −(225 ± 25 MeV)3 . Table 12 The parameters z3 () and z2 () − z1 () in the case of a() = 1 for several values of QCD and QCD

z3 ()

z2 () − z1 ()

0:30

1:00 1:10 1:20

−13:7 ± 3:1 −14:0 ± 2:2 −14:3 ± 1:6

−1:63 ± 0:60 −1:41 ± 0:32 −1:24 ± 0:19

0:35

1:00 1:10 1:20

−13:4 ± 3:6 −13:9 ± 2:5 −14:3 ± 1:8

−1:58 ± 0:69 −1:39 ± 0:37 −1:24 ± 0:22

0:40

1:00 1:10 1:20

−13:2 ± 4:0 −13:8 ± 2:8 −14:3 ± 2:0

−1:52 ± 0:80 −1:37 ± 0:43 −1:23 ± 0:25

0:45

1:00 1:10 1:20

−12:9 ± 4:5 −13:7 ± 3:2 −14:2 ± 2:2

−1:45 ± 0:92 −1:34 ± 0:49 −1:23 ± 0:28

Units of QCD and are GeV. Values of z3 () and z2 () − z1 () are scaled by a factor of 103 . The errors come from

qq S = −(225 ± 25 MeV)3 .

Now we present prediction of the several physical quantities for a() = 1 using the above bare parameters with the Wilsonian RGEs. The resultant values for (; QCD ) = (1:1; 0:4) GeV are F (0) = 73:6 ± 5:7 MeV g = 0:146 ± 0:012 GeV2

(F (0)|exp = 86:4 ± 9:7 MeV) ; (g |exp = 0:119 ± 0:001 GeV2 ) ;

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Table 13 Physical quantities predicted by the Wilsonian matching conditions and the Wilsonian RGEs for a() = 1 QCD

F (0)

g

g

a(0)

0:30

1:00 1:10 1:20

70:5 ± 6:8 66:6 ± 4:8 65:5 ± 3:3

0:144 ± 0:013 0:147 ± 0:010 0:149 ± 0:007

8:00 ± 1:19 8:73 ± 0:98 8:95 ± 0:71

2:14 ± 0:22 2:27 ± 0:18 2:31 ± 0:13

0:35

1:00 1:10 1:20

74:2 ± 7:5 70:0 ± 5:2 68:7 ± 3:7

0:143 ± 0:014 0:146 ± 0:011 0:149 ± 0:008

7:40 ± 1:12 8:09 ± 0:93 8:32 ± 0:68

2:03 ± 0:21 2:15 ± 0:17 2:20 ± 0:13

0:40

1:00 1:10 1:20

78:2 ± 8:2 73:6 ± 5:7 72:0 ± 4:0

0:143 ± 0:016 0:146 ± 0:012 0:149 ± 0:008

6:84 ± 1:06 7:49 ± 0:88 7:74 ± 0:65

1:92 ± 0:19 2:04 ± 0:16 2:09 ± 0:12

0:45

1:00 1:10 1:20

82:6 ± 8:9 77:5 ± 6:2 75:6 ± 4:4

0:143 ± 0:017 0:146 ± 0:013 0:149 ± 0:009

6:31 ± 0:99 6:93 ± 0:83 7:19 ± 0:62

1:83 ± 0:18 1:94 ± 0:15 1:99 ± 0:12

86:4 ± 9:7

0:119 ± 0:001

6:00 ± 0:01

Exp.

Units of QCD and are GeV. Unit of F (0) is MeV and that of g is GeV2 . The errors come from qq S = −(225 ± 25 MeV)3 .

g = 7:49 ± 0:88

(g |exp = 6:00 ± 0:01) ;

L9 (m ) = (7:07 ± 0:35) × 10−3 L10 (m ) = (−7:94 ± 0:84) × 10−3 a(0) = 2:04 ± 0:16 :

(L9 (m )|exp = (6:5 ± 0:6) × 10−3 ) ; (L10 (m )|exp = (−5:1 ± 0:7) × 10−3 ) ; (5.49)

We show the dependences of the results on the several choices of and QCD in Tables 13 and 14. These show that the choice a() = 1 reproduces the experimental values in reasonable agreement. To close this section, we should emphasize that the inclusion of the quadratic divergences into the RGEs was essential in the present analysis. The RGEs with logarithmic divergence alone would not be consistent with the matching to QCD. The bare parameter F () = 132–155 MeV listed in Table 5, which is derived by the matching condition (5.8), is about double of the physical value F (0) = 86:4 MeV. The logarithmic running by the 0rst term of Eq. (4.208) is not enough to change the value of F . Actually, in the present procedure with logarithmic running for (QCD ; ) = (0:4; 1:1) GeV we cannot 0nd the parameters a() and g() which reproduce F (0) = 86:4 MeV and m = 771:1 MeV. 49 49

For (QCD ; ) = (0:35; 1:0) GeV we 0nd a() and g() which reproduce F (0) = 86:4 MeV and m = 771:1 MeV as a() = 0:27 ± 0:70 ± 0:49 and g() = 4:93 ± 1:00 ± 0:69. Then, we obtain g = 0:53 GeV2 , g = 2:9, Lr9 (m ) = 15 × 10−3 and Lr10 (m ) = −30 × 10−3 . These badly disagree with experiment. Note that the parameter choice = m does not work, either.

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Table 14 Values of Gasser–Leutwyler’s parameters L9 and L10 predicted by the Wilsonian matching conditions and the Wilsonian RGEs for a() = 1 QCD

L9 (m )

L10 (m )

g =(2g F2 (0)

0:30

1:00 1:10 1:20

6:77 ± 0:38 6:71 ± 0:28 6:79 ± 0:21

−7:40 ± 0:86 −7:62 ± 0:67 −7:93 ± 0:50

1:81 ± 0:25 1:89 ± 0:19 1:94 ± 0:13

0:35

1:00 1:10 1:20

6:94 ± 0:42 6:88 ± 0:32 6:97 ± 0:23

−7:54 ± 0:96 −7:77 ± 0:75 −8:10 ± 0:56

1:76 ± 0:27 1:85 ± 0:20 1:90 ± 0:14

0:40

1:00 1:10 1:20

7:13 ± 0:47 7:07 ± 0:35 7:16 ± 0:25

−7:70 ± 1:07 −7:94 ± 0:84 −8:28 ± 0:63

1:71 ± 0:28 1:80 ± 0:21 1:86 ± 0:16

0:45

1:00 1:10 1:20

7:37 ± 0:51 7:28 ± 0:39 7:37 ± 0:28

−7:90 ± 1:19 −8:13 ± 0:93 −8:48 ± 0:70

1:66 ± 0:29 1:76 ± 0:22 1:82 ± 0:17

6:5 ± 0:6

−5:1 ± 0:7

Exp.

Units of QCD and are GeV. Values of L9 (m ) and L10 (m ) are scaled by a factor of 103 . The errors come from

qq S = −(225 ± 25 MeV)3 .

We should also stress that the above success of the Wilsonian matching is due to the existence of in the HLS. If we did not include and used the current correlators in the ChPT, we would have failed to match the e/ective 0eld theory with the underlying QCD. 5.4. Predictions for QCD with Nf = 2 As we have stressed in Section 5.1, the Wilsonian matching conditions determine the absolute values and the explicit dependence of bare parameters of the HLS on the parameters of underlying QCD such as Nc as well as Nf . Especially, the current correlators derived from the OPE has only small Nf -dependence, which implies that the bare parameters of the HLS have also small Nf -dependence. Then, the dependence of the physical quantities such as the on-shell decay constant on Nf mainly appears through the Wilsonian RGEs which do depend on Nf . In this subsection, to show how the Nf -dependences of the physical quantities appear in our framework, we consider QCD with Nf = 2. This should be regarded as a prediction for an idealized world in the in0nite strange quark mass limit (ms → ∞) of the real world. Before making a concrete analysis, let us make a rough estimation as we have done around the beginning of Section 5.2. As we stressed, the Wilsonian matching condition (5.8) determines the value of bare decay constant at the matching scale, F (). Since the dominant contribution in the right-hand side (RHS) of Eq. (5.8) is given by 1 + #s = term, the Nf -dependence of the RHS is small: The ratio F2 ()=2 has small dependence on Nf . Then, by using the matching scale as

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= 1:1 GeV, the value of F () for Nf = 2 roughly takes the same value as that for Nf = 3 as in Eq. (5.26): F2 (; Nf = 2) ∼ 3(86:4 MeV)2 ∼ (150 MeV)2 :

(5.50)

Similarly to what we have done in Eq. (5.27), we neglect the logarithmic divergence with taking a = 1 in the RGE for F2 in Eq. (4.208) to perform a crude estimate of the on-shell decay constant F (0; Nf = 2). The result is given by Nf F2 (0; Nf = 2) ∼ F2 () − 2 2(4)2

Nf 2 Nc (1 + A ) − ∼ 2 8 3 4 2 2 ∼ F2 () ∼ (120 MeV)2 ; (5.51) ∼2 4 3 where we adopted A ∼ 0:5 and Nc = 3 as in Eq. (5.27) but Nf = 2 to obtain the last line. This implies that the on-shell decay constant for Nf = 2 is about 20% bigger than that for Nf = 3 even though the bare ones have the same values. For determining all the bare parameters through the Wilsonian matching and making more precise predictions we need to determine the current correlators in the OPE. In addition to three Wilsonian matching conditions shown in Eqs. (5.8), (5.9) and (5.7), we need two inputs to determine 0ve relevant bare parameters. As we discussed above, the current correlators in the OPE have only small Nf -dependence. So, we assume that the bare parameters for Nf = 2 are the same as those obtained in Section 5.2 for Nf = 3 QCD. Then, we obtain the parameters in the low-energy region through the Wilsonian RGEs with Nf = 2 and give predictions on several physical quantities. We expect that the predictions will not be so much di/erent from the “physical quantities” obtained in the idealized QCD with Nf = 2, which can be checked by, e.g., the lattice simulation. Note that the mass m (Nf = 2) here is not an input, but an output determined from the on-shell condition in Eq. (4.217). Similarly, the on-shell decay constant F (0; Nf = 2) is also an output derived by Eq. (4.222). For de0niteness of the analysis, let us use the bare parameters determined in Nf = 3 QCD for (QCD ; ) = (0:4; 1:1) GeV. We pick up the values from Tables 6 and 7, and show them in Table 15. In Table 16 we show the physical predictions obtained from these bare parameters through the Wilsonian RGEs (4.208), (4.210), (4.211), (4.213)–(4.215) together with the on-shell condition (4.217) and relation (4.222) for the on-shell decay constant. predictions As we discussed above, the value of the decay constant, predicted as F (0; Nf = 2) = 106 MeV ;

(5.52)

is about 20% larger than that for Nf = 3 QCD, F (0; Nf = 3) = 86:4 MeV. One might think that the value F = 88 MeV estimated in Ref. [79] is the value of the pion decay constant for Nf = 2 QCD at chiral limit. However, this value is estimated from the experimental value by taking the limit of m = 0 with mK 500 MeV kept unchanged. Here we mean by Nf = 2 QCD the QCD with mu = md = 0 but ms = ∞, i.e., m = 0 but mK = ∞. In our best knowledge, there is no estimation done before for the pion decay constant in Nf = 2 QCD. But the fact that the value F = 88 MeV

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129

Table 15 Bare parameters used in the present analysis for Nf = 2 QCD. These are obtained in Nf = 3 QCD for (QCD ; ) = (0:4; 1:1) GeV (see Tables 6 and 7) F ()

a()

g()

0:145 ± 0:003

1:33 ± 0:28 ± 0:14

3:69 ± 0:13 ± 0:02

z3 ()

z2 () − z1 ()

−5:84 ± 5:78 ± 0:18

−1:79 ± 0:34 ± 0:61

Table 16 Several predictions for physical quantities in QCD with Nf = 2 done in the present analysis from the bare parameters listed in Table 15 through the Wilsonian RGEs F (0; Nf = 2)

m (Nf = 2)

0:106 ± 0:005 ± 0:002

0:719 ± 0:012 ± 0:004

g (Nf = 2)

g (Nf = 2)

a(0; Nf = 2)

0:116 ± 0:005 ± 0:002

4:30 ± 0:18 ± 0:27

1:61 ± 0:23 ± 0:13

Lr9 (m ; Nf = 2)

Lr10 (m ; Nf = 2)

9:59 ± 0:28 ± 0:27

−8:26 ± 1:92 ± 0:36

Units of F (0; Nf = 2) and m (Nf = 2) are GeV, and that of g is GeV2 . Values of Lr9 (m ) and Lr10 (m ) are scaled by a factor of 103 . Note that the 0rst and second errors correspond to those of the bare parameters in Table 15.

for m = 0 but mK 500 MeV is slightly larger than F (0; Nf = 3) = 86:4 MeV for m = mK = 0 indicates that increase of 20% may be possible when we change the value of ms (thus mK ) from zero to in0nity. On the other hand, the mass is predicted as m (Nf = 2) = 719 MeV ;

(5.53)

which is about 10% smaller than the experimental value m (Nf =3)=771:1 MeV. This is mainly due to the smallness of the HLS gauge coupling g(m ): The present analysis provides g(m ; Nf = 2) = 5:33 ± 0:53 ± 0:10 to be compared with g(m ; Nf = 3) = 6:27 ± 0:69 ± 0:11 in Table 8. Accordingly, the absolute values of L9 and L10 becomes larger since their main parts are determined by 1=g2 (m ). Finally, the predicted value of a(0) shows that there exists the deviation from 2, which implies that the vector dominance (VD) is violated in Nf = 2 QCD. This also implies that the VD in the real world (QCD with Nf = 3) can be realized only accidentally (see Sections 4.10 and 6.3.4).

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5.5. Spectral function sum rules In this subsection we study the spectral function sum rules (the Weinberg sum rules and Das–Mathur–Okubo sum rule), which are related to the vector and axialvector current correlators. The spectral function sum rules are given by

∞ ds [V (s) − A (s)] = −4LS10 ; (5.54) s 0

∞ ds[V (s) − A (s)] = F2 ; (5.55) 0

0

∞

ds s[V (s) − A (s)] = 0 ;

(5.56)

where LS10 is a constant which corresponds to the so-called S parameter in the electroweak theory [6,138,114,156,157] as LS10 → −S=(16). The relations in Eqs. (5.55) and (5.56) are called the Weinberg’s 0rst and second sum rules [184], respectively, and we call the relation in Eq. (5.54) Das–Mathur–Okubo (DMO) sum rule [62]. In the above expressions, V (s) and A (s) are the spin 1 parts of the spectral functions of the vector and axialvector currents. These spectral functions are de0ned by

1 d 4 x eipx 0|J$a (x)Jb (0)|0 = ab (p$ p − g$ p2 )V (p2 ) ; (5.57) 2

1 a b d 4 x eipx 0|J5$ (x)J5 (0)|0 = ab p$ p 0A (p2 ) + ab (p$ p − g$ p2 )A (p2 ) ; (5.58) 2 where 0A (p2 ) is the spin 0 part. By using these spectral functions, the g$ -term and the p$ p -term of the VV − AA current correlator are expressed as

s{V (s) − A (s)} 1 (1) 2 ; (5.59) ds TV −A (−p ) = 2 −p s − p2 − ij

V (s) − A (s) − 0A (s) (2) 2 ; (5.60) TV −A (−p ) = ds s − p2 − ij where TV(1)−A and TV(2)−A are related to the VV − AA current correlator as

a b (x)J5 (0)|0 ] i d 4 x eipx [0|TJ$a (x)Jb (0)|0 − 0|TJ5$ =ab [p$ p TV(2)−A (−p2 ) − g$ p2 TV(1)−A (−p2 )] :

(5.61)

Note that both TV(1)−A and TV(2)−A agree with TV − TA de0ned in Eq. (5.1) when the current conservation is satis0ed in the chiral limit (massless current quark): TV(1)−A (−q2 ) = TV(2)−A (−p2 ) = TV (−p2 ) − TA (−p2 )

(for mq = 0) :

(5.62)

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For the convergence of the above sum rules a crucial role is played by the asymptotic behavior of the spectral functions which is rephrased by the requirement for the high energy behavior of the VV − AA current correlator: The convergence of the sum rules in Eqs. (5.54), (5.55) and (5.56), respectively, requires that the VV − AA current correlator must satisfy TV(1)−A (Q2 ) = TV (Q2 ) − TA (Q2 ) → 0 ;

(5.63)

Q2 TV(2)−A (Q2 ) = Q2 [TV (Q2 ) − TA (Q2 )] → 0 ;

(5.64)

Q4 TV(1)−A (Q2 ) = Q4 [TV (Q2 ) − TA (Q2 )] → 0 ;

(5.65)

Q2 →∞

Q2 →∞ Q2 →∞

where Q2 = −p2 . 50 It should be noticed that the VV − AA current correlator obtained by the OPE in QCD satis0es in the chiral limit all the above convergence conditions as [see Eqs. (5.5) and (5.6)] TV(QCD) (Q2 ) − TA(QCD) (Q2 ) → − Q2 →∞

4(Nc2 − 1) #s qq S 2 : Nc2 Q6

(5.66)

Provided the above convergence conditions, the DMO sum rule and the 0rst Weinberg sum rule are rewritten as the following relations in the low-energy region: d 2 2 2 2 [Q TV (Q ) − Q TA (Q )] = −4LS10 ; (5.67) 2 dQ Q2 =0 [ − Q2 TV (Q2 ) + Q2 TA (Q2 )]|Q2 =0 = F2 :

(5.68)

It should be noticed that there is an infrared divergence in Eq. (5.67) coming from the -loop contribution. To regularize the infrared divergence we introduce the mass when we consider the DMO sum rule. 51 In such a case, the constant LS10 is related to the axialvector form factor FA of ± → ‘+ studied in Section 2.9 and the charge radius of pion r 2 V studied in Section 2.8 as [62] F FA F2 ± − 4LS10 = − √ + r 2 V ; 3 2m± which is related to the ChPT parameter Lr10 ($) in Section 2.9 as 52

Nf m2 − 4LS10 = −4L10 ($) − ln + 1 : 6(4)2 $2

(5.69)

(5.70)

Let us show how the DMO sum rule and the 0rst and second Weinberg’s sum rules are satis0ed in the present approach. As we have shown above, the spectral function sum rules under consideration are equivalent to the combination of the convergence conditions (5.63)–(5.65) and the low-energy relations (5.67) and (5.68). In the following, therefore, we consider only the current correlators. 50 When we wrote the dispersive form as in Eq. (5.59) with no subtraction, we implicitly assumed the converge condition in Eq. (5.63). Then, the form in Eq. (5.59) automatically satis0es Eq. (5.63). 51 As can be seen in, e.g., Refs. [33,197] introduction of the mass, or equivalently the current quark mass, changes the higher energy behavior of the current correlators in such a way that the convergence conditions in Eqs. (5.64) and (5.65) are not satis0ed while that in Eq. (5.63) is still satis0ed. Thus, we do not include the mass when we consider the 0rst and second Weinberg’s sum rules, while for the DMO sum rule we include it as an infrared regulator. 52 We can check the validity of Eq. (5.70) for Nf = 3, especially the second term in the square bracket by substituting ± the expression of FA in Eq. (2.69) and that of r 2 V in Eq. (2.61) with mK = m into Eq. (5.69).

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In the present approach we switch the theory from the HLS to QCD at the matching scale . In other words, in the energy region below we use the HLS, while in the energy region above we use QCD. Then, the vector and axialvector current correlators may be expressed as 53 (HLS) (QCD) 2 2 2 2 TV; A (Q2 ) = :(2 − Q2 )TV; A (Q ) + :(Q − )TV; A (Q ) ;

(5.71)

(HLS) 2 where TV; A (Q ) are the correlators obtained by the HLS in the energy region below and (QCD) TV; A (Q2 ) are those obtained by QCD in the energy region above . Thus we have

TV (Q2 ) − TA (Q2 ) → TV(QCD) (Q2 ) − TA(QCD) (Q2 ) → − Q2 →∞

Q2 →∞

S 2 4(Nc2 − 1) #s qq : Nc2 Q6

(5.72)

This implies that the current correlators in the present approach satisfy all the convergence conditions (5.63)–(5.65). The fact that the VV − AA current correlator satis0es the convergence condition in Eq. (5.65) already implies that the second Weinberg’s sum rule is satis0ed in the present approach. The next issue for showing the DMO sum rule and the 0rst Weinberg’s sum rule is to check whether or not the low-energy relations (5.67) and (5.68) are satis0ed. For this purpose we consider the vector and axialvector current correlators in the HLS. In the HLS at one loop the vector and axialvector current correlators are given by [see also Eqs. (4.220) and (4.224)] TV(HLS) (−p2 ) = TA(HLS) (−p2 ) =

TVS (p2 ) [ − TVT (p2 ) − 2TVT (p2 )] − T T (p2 ) ; + p2 TVT (p2 )

TVS (p2 ) S

T⊥ (0) T − T˜ S⊥ (p2 ) − T⊥ (p2 ) ; −p2

(5.73) (5.74)

where S (0) TS (p2 ) − T⊥ : T˜ S⊥ (p2 ) = ⊥ 2 p

(5.75)

S (0) = F2 (0), the low-energy relation (5.68) is satis0ed, which together with the converSince T⊥ gence condition in Eq. (5.64) implies that the 0rst Weinberg’s sum rule is actually satis0ed in the present approach. By using Eqs. (4.228) and (4.246) together with Eq. (4.237), the DMO sum rule Eq. (5.67) takes the form:

− 4LS10 53

1 1 Nf m2 S − 2z (m ) − 2z (m ) + T (0) + 2z (m ) − ln ; 3 1 2 ⊥ g2 (m ) 6 (4)2 m2

(5.76)

More precisely, our matching conditions Eqs. (5.7)–(5.9) read: lim

[TV (Q2 ) − TA (Q2 )] =

lim

d d TV; A (Q2 ) = lim TV; A (Q2 ) : 2 dQ2 Q2 →2 +j dQ

Q2 →2 −j

Q2 →2 −j

lim

[TV (Q2 ) − TA (Q2 )] ;

Q2 →2 +j

(HLS) 2 2 Note that a low-energy expansion of the TV; A (Q ) is in positive powers of Q , while the high-energy expansion or the (QCD) OPE of the TV; A (Q2 ) is in negative power of Q2 . Our matching condition thus is a best compromise between these two with di/erent Q2 behaviors.

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where we put the mass m to regularize the infrared divergence. By putting the HLS parameters determined in Section 5.3.1 into the right-hand side of Eq. (5.76), the value of LS10 for (; QCD ) = (1:1; 0:4) GeV is estimated as LS10 = (−8:5 ± 2:5 ± 0:6) × 10−3 ;

(5.77)

where the 0rst error comes from the error of the input value of F (0); F (0) = 86:4 ± 9:7 MeV, and the second error from that of qq , S qq S 1 GeV = (−225 ± 25 MeV)3 . This is to be compared with the experimental value LS10 |exp = (−7:0 ± 0:2) × 10−3 ;

(5.78) ±

obtained by substituting the experimental values of FA given in Eq. (2.70) and r 2 V = 0:455 ± 0:005 fm2 from the most recent data [72] in Table 2 into Eq. (5.69). Here, let us consider the pole-saturated form of the sum rules which are usually saturated by , and a1 . When we assume that the vector and the axialvector current correlators are saturated by , and a1 , they are expressed as TV(pole) (−p2 ) =

(g =m )2 ; m2 − p2 − ij

TA(pole) (−p2 ) =

(ga1 =ma1 )2 F2 + ; −p2 − ij m2a1 − p2 − ij

(5.79)

where F , m , ma1 , g and ga1 are the parameters at the on-shell of corresponding particles. Note that the above forms written by the on-shell parameters are valid only around the on-shell of the relevant particles, and that we have no guarantee to use the same forms in the o/-shell region, especially in the high-energy region. Nevertheless, as customarily done, we may assume that the above forms are valid even in the high-energy region. In such a case, the above correlators must satisfy the convergence conditions in Eqs. (5.63)–(5.65) as well as the low-energy relations in Eqs. (5.67) and (5.68). As we can see easily, the above correlators satisfy the convergence condition (5.63) corresponding to the DMO sum rule. On the other hand, the convergence conditions (5.64) and (5.65) corresponding to the 0rst and second Weinberg’s sum rule require that the parameters in Eq. (5.79) must satisfy g2 ga21 2 = F + ; m2 m2a1

(5.80)

g2 = ga21 :

(5.81)

Eq. (5.80) implies that the low-energy relation (5.68) corresponding to the 0rst Weinberg’s sum rule is already satis0ed. The low-energy relation (5.67) corresponding to the DMO sum rule is satis0ed as − 4LS10 =

g2 ga21 1 Nf m2 − − ln ; m4 m4a1 6 (4)2 m2

(5.82)

where we added the last term to include the possible contribution from the loop with the infrared regularization. Eqs. (5.82), (5.80) and (5.81) are the pole saturated forms of the DMO sum rule and the 0rst and the second Weinberg’s sum rules.

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One might think that the spectral function sum rules in Eqs. (5.54)–(5.56) always lead to the above relations in Eqs. (5.82), (5.80) and (5.81), and hence the existence of a1 meson is inevitable. However, it is not true: It is merely the peculiarity of the assumption of the pole saturation. In our approach, on the other hand, we have demonstrated that the sum rules are saturated in a di/erent manner without a1 meson which is heavier than the scale . Then, it does not make sense to consider the above relations in Eqs. (5.82), (5.80) and (5.81) in the framework of the present approach. Nevertheless, it may be worth showing how the a1 contribution in the pole saturated form is numerically reproduced in the present approach. Using the de0nition of g given in Eq. (5.35) together with the de0nition of the on-shell mass m2 = g2 (m )F2 (m ), we obtain g2 1 2 (5.83) − 2z3 (m ) ; 4 m g (m ) where we neglected the higher order corrections. Comparing Eq. (5.82) with Eq. (5.76) and using Eq. (5.83), we see the following correspondence: ga21 S ⇔ 2[z1 (m ) − z2 (m )] − T⊥ (0) : (5.84) 4 ma 1 This implies that the a1 contribution in the pole saturated form of the DMO sum rule in Eq. (5.82) S is numerically imitated by especially the – loop contribution [177] expressed by T⊥ (0) in the present approach. In a similar way, the – loop contribution does yield additional contribution to the axialvector correlator, as shown by T˜ S⊥ (p2 ) in Eq. (5.74). This actually gives an imaginary part (i.e., the additional contribution to the spectral function) above the – threshold and hence mimic the a1 pole e/ects in the 0rst and second Weinberg’s sum rules. As we have shown above, while the current correlators obtained in our approach within the framework of the HLS do satisfy the spectral function sum rules, the pole saturated form of the 0rst and second Weinberg’s sum rules are not generally reproduced as it stands since a1 is not explicitly included in our approach. Nevertheless, there is a special limit where the pole saturated forms without a1 contribution, i.e., g2 = F2 (0) ; (5.85) m2 (5.86) g2 = 0 ; are well reproduced. This in fact occurs at the limit of the vector manifestation (VM) which will be studied in detail in Section 6. In the VM, the chiral symmetry is restored at the critical point by the massless degenerate and the as the chiral partner, which is characterized by [see Eq. (6.2) as well as Eq. (5.44)] F2 (0) → 0;

m2 → m2 = 0;

a(0) = F2 (m )=F2 (0) → 1 ;

(5.87)

where F (m ) is the decay constant of (longitudinal ) at on-shell. As we will show in Section 6, the VM is realized within the framework of the HLS due to the fact that, at the chiral restoration point, the bare parameters of the HLS determined from the Wilsonian matching satisfy the VM conditions given in Eqs. (6.11)–(6.14) which lead to the following condition for the parameter g(m ) at the on-shell of [see Eq. (6.17)]: g(m ) → 0 :

(5.88)

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Using the expression of g in terms of the parameters of the HLS given in Eq. (5.35) and the on-shell condition m2 = g2 (m )F2 (m ) together with Eqs. (5.88) and (5.87) we obtain g2 F2 (m ) 1 = [1 − 2g2 (m )z3 (m )] → 1 : m2 F2 (0) F2 (0)

(5.89)

This implies that the pole saturated form of the 0rst Weinberg’s sum rule without a1 given in Eq. (5.88) is actually satis0ed at the VM limit. Furthermore, Eq. (5.88) already implies that the second Weinberg’s sum rule is satis0ed even without a1 contribution at the VM limit: g2 = g2 (m )[1 − 2g2 (m )z3 (m )] → 0 :

(5.90)

6. Vector manifestation Chiral symmetry restoration (Wigner realization of chiral symmetry) is an outstanding phenomenon expected in QCD under extreme conditions such as the 0nite temperature and/or density (for reviews, see, e.g., Refs. [43,45,109,111,160,162,194]), the large Nf (3 ¡ Nf ¡ 33=2), Nf being the number of massless >avors (see, e.g. Refs. [12,14,41,61,117–122,148]), etc. Conventional picture of the chiral symmetry restoration is based on the linear sigma model where the scalar meson (“sigma” meson) denoted by S becomes massless degenerate with the pion as the chiral partner: F2 (0) → 0;

m2S → m2 = 0 :

(6.1)

This we shall call “GL manifestation” after the e/ective theory of Ginzburg–Landau or Gell–Mann– Levy. However, the GL manifestation is not a unique way where the Wigner realization manifests itself. Recently the present authors [106] proposed “vector manifestation (VM)” as a novel manifestation of Wigner realization of chiral symmetry where the vector meson becomes massless at the chiral phase transition point. Accordingly, the (longitudinal) becomes the chiral partner of the NG boson . The VM is characterized by F2 (0) → 0;

m2 → m2 = 0;

F2 (m )=F2 (0) → 1 ;

(6.2)

where F (m ) is the decay constant of (longitudinal ) at on-shell. Here we should stress that the power counting rule in our derivative expansion developed in Section 4.1, which presumes mass is conceptually small in the same sense as mass, is now literally (not just conceptually) operative near the VM phase transition, although it is not a priori justi0ed for the case Nf = 3 where m is actually not very small, except that it happened to work as demonstrated in Section 5. In this section we discuss the VM of chiral symmetry, based on the HLS model at one loop developed in the previous sections: In Section 6.1 we 0rst formulate in Section 6.1.1 what we call “VM conditions”, Eqs. (6.11)– (6.14), a part of which coincides with the Georgi’s “vector limit” [85,86]. The VM conditions are necessary conditions of the Wigner realization of chiral symmetry of QCD in terms of the HLS parameters as a direct consequence of the Wilsonian matching of the HLS with the underlying QCD at the matching scale . We then argue that we have the chiral restoration F2 (0) → 0 through the dynamics of the HLS model itself in a way already discussed in Section 4.10, once the VM

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conditions are imposed on the bare parameters of the HLS model for a particular value of F () and/or Nf such that X () ≡ (Nf 2 =2(4)2 )=F2 () → 1, where X ($) was de0ned by Eq. (4.254). Then we show that the VM conditions in fact lead to VM. We compare the VM with the conventional manifestation, i.e., GL manifestation in Section 6.1.2: we demonstrate that the GL manifestation a^ la linear sigma model does not satisfy the requirement on the current correlators from the Wilsonian matching (i.e., VM conditions), and hence is excluded by the Wilsonian matching as a candidate for the chiral restoration of QCD. In Section 6.1.3 we discuss the “conformal phase transition” [148] as an example of non-GL manifestation having the essential-singularity-type scaling. In Section 6.1.4 we distinguish our VM as a Wigner realization from a similar but essentially di/erent concept, the “vector realization” [85,86], which was claimed as a new realization, neither Wigner nor NG realization. In Section 6.1.5 we emphasize that the VM makes sense only as a limit of the bare parameters approaching the values of VM conditions (never does the “vector realization” even as a limit). In Section 6.2, as an illustration of VM we shall discuss the chiral restoration in the large Nf QCD: we 0rst review the arguments on chiral restoration in the large Nf QCD in terms of the QCD language, i.e., qq S → 0. It is noted that the conformal phase transition was observed also in the chiral restoration of the large Nf QCD in the (improved) ladder approximation [14]. In Section 6.3 we show that the chiral restoration in the large Nf QCD in fact takes place also in the HLS model, F2 (0) → 0, and so does the VM, when we tune in a concrete manner the bare parameters to satisfy the above condition X () ≡ (Nf 2 =2(4)2 )=F2 () → 1. In Section 6.3.1 we determine by this the critical number of >avors Nf = Nfcrit 5 above which the chiral symmetry is restored, which is in rough agreement with the recent lattice simulations 6 ¡ Nfcrit ¡ 7 [118]. The critical behaviors of the parameters in the large Nf QCD are studied in Section 6.3.2. Full Nf -dependences of the parameters are shown in Section 6.3.3 by using a simple ansatz. In Section 6.3.4 we argue, following Ref. [107], that the vector dominance is badly violated near the critical point in the large Nf QCD. Finally, in Section 6.4 we explain the proposal of Ref. [104] that the HLS in the broken phase of chiral symmetry is dual to QCD in the sense of Seiberg duality [170]. 6.1. Vector manifestation (VM) of chiral symmetry restoration 6.1.1. Formulation of the VM The essence of VM stems from the new matching of the EFT with QCD (Wilsonian matching) proposed by Ref. [105] [see Section 5] in which bare parameters of the EFT are determined by matching the current correlators in the EFT with those obtained by the OPE in QCD, based on the RGE in the Wilsonian sense including the quadratic divergence [104] [see Section 4]. Several physical quantities for and were predicted by the Wilsonian matching in the framework of the HLS model [21,24] as the EFT, in excellent agreement with the experiments for Nf = 3, where Nf is the number of massless Davors [105]. This encourages us to perform the analysis for other situations such as larger Nf and 0nite temperature and/or density up to near the critical point, based on the Wilsonian matching. The chiral symmetry restoration in Wigner realization should be characterized by F (0) = 0

(6.3)

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(see also discussions in Section 6.1.4) and the equality of the vector and axialvector current correlators in the underlying QCD: TV (Q2 ) = TA (Q2 ) ;

(6.4)

which is in accord with qq S = 0 in Eqs. (5.5) and (5.6). On the other hand, the same current correlators are described in terms of the HLS model for energy lower than the cuto/ : When we approach to the critical point from the broken phase (NG phase), the axialvector current correlator is still dominated by the massless as the NG boson, while the vector current correlator is by the massive . In such a case, there exists a scale around which the current correlators are well described by the forms given in Eqs. (5.2) and (5.3): TA(HLS) (Q2 ) =

F2 () − 2z2 () ; Q2

(6.5)

TV(HLS) (Q2 ) =

F2 () [1 − 2g2 ()z3 ()] − 2z1 () ; M2 () + Q2

(6.6)

where M2 () ≡ g2 ()F2 () is the bare mass parameter [see Eq. (5.4)]. Then, through the Wilsonian matching discussed in Section 5, we determine the bare parameters of the HLS. At the critical point the quark condensate vanishes, qq S → 0, while the gluonic condensate (#s =)G$ G $ is independent of the renormalization point of QCD and hence it is expected not to vanish. Then Eq. (5.8) reads Nc 2 2 crit 2 F () → (F ) ≡ · 2(1 + crit A ) = 0; 3 4 crit S =0 = A ≡ A |qq

3(Nc2 − 1) #s 22 (#s =)G$ G $ + ¿0 8Nc Nc 4

(1) ;

(6.7)

implying that matching with QCD dictates F2 () = 0

(6.8)

even at the critical point [106] where F2 (0) = 0 :

(6.9)

One might think that this is somewhat strange. However, as we have already discussed in Sections 4.5.2 and 4.10, we have a possibility [104] that the order parameter can become zero F (0) → 0, even when F () = 0, where F () is not an order parameter but just a parameter of the bare HLS Lagrangian de0ned at the cuto/ where the matching with QCD is made. Let us obtain further constraints on other bare parameters of the HLS through the Wilsonian matching for the currents correlators. The constraints on other parameters de0ned at come from the fact that TA(QCD) and TV(QCD) in Eqs. (5.5) and (5.6) agree with each other for any value of Q2 when the chiral symmetry is restored with qq S → 0. Thus, we require that TA(HLS) and TV(HLS) in

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Eqs. (6.5) and (6.8) agree with each other for any value of Q2 (near 2 ). 54 Under the condition Eq. (6.8), this agreement is satis0ed only if the following conditions are met: M2 () ≡ g2 ()F2 () → 0;

F2 () → F2 () = 0;

z1 () − z2 () → 0 ;

(6.10)

or g() → 0 ; a() =

(6.11)

F2 () →1 ; F2 ()

z1 () − z2 () → 0 ; F2 () → (Fcrit )2 =

Nc 3

(6.12) (6.13)

4

2

· 2(1 + crit A ) = 0 :

(6.14)

These conditions, may be called “VM conditions”, follow solely from the requirement of the equality of the vector and axialvector currents correlators (and the Wilsonian matching) without explicit requirement of Eq. (6.3), and are actually a precise expression of the VM in terms of the bare HLS parameters for the Wigner realization in QCD [106]. Note that the values in Eqs. (6.11) and (6.12) agree with the values in the Georgi’s vector limit [85,86]. Once the bare HLS parameters satisfy the VM conditions, Eqs. (6.11)–(6.14), the RGE for F2 leads to Eq. (4.253), N f 2 Nf 2 crit 2 F2 (0) = F2 () − → (F ) − ; (6.15) 2(4)2 2(4)2 which implies that we can have F2 (0) → 0

(6.16)

by tuning the bare parameters Nf and/or F2 () (which explicitly depends on Nc ) in such a way that X () ≡ [Nf 2 =2(4)2 ]=F2 () → 1. Then the chiral restoration F2 (0) → 0 is actually derived within the dynamics of the HLS model itself solely from the requirement of the Wilsonian matching. (We shall discuss a concrete way of tuning the bare parameters in the case of large Nf QCD in Section 6.3.) One may wonder what would happen if we tune the HLS parameters such as Nf so as to keep X () = 1 even when the bare parameters obey the VM conditions: in such a case the underlying QCD gives a chiral restoration, qq S = 0, while the EFT would have an NG boson pole coupled to the axialvector current with the strength of a pole residue F2 (0) = 0! This is similar to the Georgi’s “Vector Realization” [85,86]. We shall discuss in details in Section 6.1.4 that the Note that chiral restoration requires equality of TA(HLS) and TV(HLS) for any Q2 (even without referring to QCD), while Eqs. (6.5) and (6.6) are valid only for Q2 ∼ 2 . See the discussions below Eqs. (5.5) and (5.6). For instance, the forms in Eqs. (6.5) and (6.6) might be changed for Q2 ¡ 2 by the corrections to TV and TA from and/or loop e/ects which, however, are of higher order in our power counting rule developed in Section 4.1 and hence can be neglected. Note that the counting rule actually becomes precise near the VM limit satisfying the VM conditions. Also note that the VM limit is the 0xed point and hence the “pole-saturated forms” of Eqs. (6.5) and (6.6) must be equal for any Q2 , once the VM conditions are satis0ed at Q2 ∼ 2 : Namely, other possible e/ects if any should be equal to each other at the VM limit and hence would not a/ect our arguments. 54

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“Vector Realization” is in contradiction with the Ward–Takahashi identity for the chiral symmetry and also produces a fake symmetry larger than the underlying QCD and hence is impossible. So the parameters of HLS model must choose a choice such that X () → 1 or F2 (0) → 0. Now that we have shown the Wigner realization in the HLS model, we can show that the VM conditions actually lead to the VM characterized by Eq. (6.2): First note that since the values in Eqs. (6.11)–(6.13) coincide with those at the 0xed points of the RGEs [see Eqs. (4.212) and (4.216)], the parameters remains the same for any scale, and hence even at on-shell point: g(m ) → 0 ;

(6.17)

a(m ) → 1 ;

(6.18)

z1 (m ) − z2 (m ) → 0 ;

(6.19)

where m is determined from the on-shell condition in Eq. (4.217): m2 = a(m )g2 (m )F2 (m ) :

(6.20)

Then, the condition in Eq. (6.17) together with the above on-shell condition immediately leads to m2 → 0 :

(6.21)

Eq. (6.18) is rewritten as F2 (m )=F2 (m ) → 1, and Eq. (6.21) implies F2 (m ) → F2 (0). Thus, F2 (m )=F2 (0) → 1 ;

(6.22)

namely, the pole residues of and become identical. Then the VM de0ned by Eq. (6.2) does follow. Note that we have used only the requirement of Wigner realization in QCD through the Wilsonian matching and arrived uniquely at VM but not GL manifestation aJ la linear sigma model. The crucial ingredient to exclude the GL manifestation as a chiral restoration in QCD was the Wilsonian matching, particularly Eq. (6.8). We shall return to this point later in Section 6.1.2. Actually, the VM conditions with X () → 1 are nothing but a limit of bare parameters approaching a particular 0xed point (what we called “VM limit”) (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0) in Eq. (4.260) which was extensively discussed in Section 4.10. Namely, through the VM conditions the QCD singles out just one ?xed point (as a limit) out of otherwise allowed wide phase boundary surface of HLS model which is given by the collection of the RG >ows entering points on the line speci0ed by Eqs. (4.259) and (4.265). Now, does it make sense that Lorentz scalar and Lorentz vector are the chiral partner? It is crucial that only the longitudinal component of becomes a chiral partner of , while the transverse decouples. When the VM occurs, both the axialvector and vector current correlators in Eqs. (6.5) and (6.6) take the form [106] TA(HLS) (Q2 ) =

F2 () F2 () − 2z () = − 2z1 () = TV(HLS) (Q2 ) : 2 Q2 Q2

(6.23)

For the axialvector current correlator, the 0rst term F2 ()=Q2 (=F2 ()=Q2 ) comes from the exchange contribution, while for the vector current correlator it can be easily understood as the (would-be NG boson absorbed into )-exchange contribution in the RC -like gauge. Thus only the

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longitudinal couples to the vector current, and the transverse with the helicity ±1 is decoupled from it [106]. This can be also seen in the unitary gauge as follows: Let us start with the expression of the vector current correlator in the chiral broken phase in the unitary gauge of : p$ p (gF2 )2 (HLS) T(V )$ (p) = − 2 g$ − + g$ F2 ; (6.24) m − p 2 m2 where p$ =(p0 ; p ˜ ) and we have neglected higher order z3 and z1 terms for simplicity. The polarization vector for the longitudinal is given by p ˜ 1 (0) ; (6.25) A$ (P) = |˜ p|; E m |˜ p| where P$ ≡ (E; p ˜ ) with E = |˜ p|2 + m2 . It is given for the transverse by A$(±) (P) = (0; e(˜±) (˜ p)) ;

(6.26)

˜ (˜ p) satisfy e(˜±) (˜ p) · p ˜ = 0, e(+) p) · e(˜−) (˜ p) = 0 and e(˜±) (˜ p) · e(˜±) (˜ p) = 1. Using a relation where e(˜±) (˜ P$ P ; (6.27) A$(l) (P)A(l) (P) = − g$ − 2 m l=±;0

we can rewrite Eq. (6.24) into (gF 2 )2 (HLS) 2 T(V A$(l) (P)A(l) (P) 2 2 )$ (p) = g$ F + m − p l=±

+ A$(0) (P)A(0) (P)

(gF2 )2 F2 + (p p − P P ) : $ $ m2 − p2 m2 − p2

(6.28)

Let us consider VM such that (g; F ) → (0; F ). We can easily show gA$(±) → 0

(6.29)

from Eq. (6.26). This implies that the transverse components of decouple from the vector current. On the other hand, Eq. (6.25) leads to 1 1 gA$(0) → (|˜ p|; p ˜) = P$ ; (6.30) F F where we used E → |˜ p| as m → 0. Eq. (6.30) implies that the longitudinal component of does couple to the vector current. The resultant expression of the vector current is given by (HLS) 2 T(V )$ (p) = (p$ p − p g$ )

F2 ; −p2

(6.31)

which agrees with the axial vector current correlator as it should. 6.1.2. VM vs. GL (Ginzburg–Landau/Gell–Mann–Levy) manifestation The crucial ingredient of the Wilsonian matching is the quadratic divergence of HLS model which yields the quadratic running of (square of) the decay constant F2 ($) [104], where $ is the

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renormalization point. Then the contribution to the axialvector current correlator at $ = 0 persists, F ($) = 0, even at the critical point where F (0) = 0. Thus the only possibility for the equality TA = TV to hold at any $ = 0 is that the contribution to the vector current correlator also persists at the critical point in such a way that yields a massless pole with the current coupling equal to that of , i.e., the VM occurs: the chiral restoration is accompanied by degenerate massless and (longitudinal) (the would-be NG boson ). 55 On the contrary, the scalar meson in the linear sigma model does not contribute to TV and hence the GL manifestation aK la linear sigma model (without ) is simply ruled out by Eq. (6.8): The Wilsonian matching with QCD de0nitely favors VM rather than GL manifestation. Let us discuss the di/erence between the VM and GL manifestation in terms of the chiral representation of the mesons by extending the analyses done in Refs. [87,186] for two >avor QCD. Since we are approaching the chiral restoration point only from the broken phase where the chiral symmetry is realized only nonlinearly, it does not make sense to discuss the chiral representation of such a spontaneously broken symmetry. One might suspect that in the HLS model having the linearlized symmetry Gglobal × Hlocal , the is an adjoint representation of the gauge symmetry Hlocal and is a singlet of the chiral symmetry Gglobal . However, the Gglobal × Hlocal is actually spontaneously broken down to H , which is a diagonal subgroup of Hglobal (⊂ Gglobal ) and Hlocal , and hence the is no longer subject to the linear representation. Then we need a tool to formulate the linear representation of the chiral algebra even in the broken phase, namely the classi0cation algebra valid even in the broken phase, in such a way that it smoothly moves over to the original chiral algebra as we go over to the symmetric phase. Following Ref. [186], we de0ne the axialvector coupling matrix Xa (6) (an analog of the gA for the nucleon matrix) by giving the matrix elements at zero invariant momentum transfer of the axialvector current between states with collinear momenta as 56 + ˜q6 L|J5a (0)|˜ p6# = 2p+ 66 [Xa (6)]L# ; (6.32) √ + 0 3 where J5a = (J5a + J5a )= 2, and # and L are one-particle states with collinear momentum p ˜ ≡ + 1 2 (p ; p ; p ) and ˜q ≡ (q+ ; q1 ; q2 ) such that p+ = q+ , 6 and 6 are their helicities. It was stressed [186] that the de0nition of the axialvector couplings in Eq. (6.32) can be used for particles of arbitrary spin, and in arbitrary collinear reference frames, including both the frames in which |# is at rest and in which it moves with in0nite momentum: The matrix Xa (6) is independent of the reference frame. Note that the Xa (6) matrix does not contain the pole term which would behave as (p+ − q+ )=[(p − q)2 − m2 ] and hence be zero for kinematical reason, p+ = q+ , even in the chiral limit of m2 → 0. 55

The transverse is decoupled from the current correlator in the limit approaching the critical point, as we discussed around Eq. (6.29). Note that when the theory is put exactly on the critical point, then not only the transverse but also the whole light spectrum including the and the longitudinal would disappear as we shall discuss in Sections 6.1.3, 6.1.4 and 6.2. The e/ective 0eld theory based on the light composite spectrum would break down at the exact critical point. 56 Note that we adopted the invariant normalization for the state:

˜q6 L|˜ p6# = (2)3 2p+ (˜q − p ˜) ; which is di/erent from the one used in Ref. [186]. Furthermore, the current in this expression is half of the current used in Ref. [186].

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As was done for Nf = 2 in Ref. [186], considering the forward scattering process a + #(6) → b + L(6 ) and requiring the cancellation of the terms in the t-channel, we obtain [Xa (6); Xb (6)] = ifabc Tc ;

(6.33)

where Tc is the generator of SU(Nf )V and fabc is the structure constant. This is nothing but the algebraization of the Adler–Weisberger sum rule [1,191] and the basis of the good-old-days classi0cation of the hadrons by the chiral algebra [87,186] or the “mended symmetry” [187]. It should be noticed that Eq. (6.33) tells us that the one-particle states of any given helicity must be assembled into representations of chiral SU(Nf )L × SU(Nf )R . Furthermore, since Eq. (6.33) does not give any relations among the states with di/erent helicities, those states can generally belong to the di/erent representations even though they form a single particle such as the longitudinal (6 = 0) and the transverse (6 = ±1). Thus, the notion of the chiral partners can be considered separately for each helicity. Here we should note that the above axialvector% coupling matrix Xa (6) can be equivalently de0ned + through the light-front (LF) axial charge Qˆ 5a ≡ d x− d x1 d x2 J5a (x) as 3 + 3 p6# = (2) 2p (˜ p − ˜q) 66 [Xa (6)]L# : (6.34) ˜q6 L|Qˆ 5a |˜ The LF axial charge Qˆ 5a does not contain the pole term for the same reason as the absence of pole contribution in the Xa (6) matrix and is well de0ned even in the chiral limit in the broken phase in such a way that the vacuum is singlet under the chiral transformation with Qˆ 5a , (6.35) Qˆ 5a |0 = 0 ; whereas the ordinary axial charge Q5a is not well de0ned due to the presence of the pole, or usually phrased as Q5a |0 = 0. However, due to the very absence of the pole term, Qˆ 5a is not conserved even in the chiral limit m2 → 0 in the broken phase: d (6.36) i + Qˆ 5a = [Qˆ 5a ; P − ] = 0 ; dx √ in sharp contrast√ to the conservation of Q5a , where x+ = (x0 + x3 )= 2 is the LF time and P − = (P 0 − P 3 )= 2 is the LF Hamiltonian. Then it does not commute with the (mass)2 operator M 2 = 2P + P − − (P 1 )2 − (P 2 )2 : (6.37) [Qˆ 5a ; M 2 ] = 0 : This implies that the mass eigenstates are in general admixtures of the representations of the chiral algebra (LF chiral algebra) which is formed by the LF axial charge Qˆ 5a together with the LF vector charge Qˆ a . This is nothing but the representation mixing in the saturation scheme [87,186,187] of the celebrated Adler–Weisberger sum rule which is actually a physical manifestation of the LF chiral algebra. When the symmetry is restored with vanishing pole, the LF axial charge agrees with the ordinary axial charge, and then the representations of the algebra with Qˆ 5a agree with the ones under the ordinary axial charge. (For details of the LF charge algebra, see Ref. [201].) The same is of course true for the algebra formed by the Xa (6) matrix directly related to Qˆ 5a through Eq. (6.34). In the broken phase of chiral symmetry, the Hamiltonian (or (mass)2 ) matrix 2 M#L de0ned by the matrix elements of the Hamiltonian ((mass)2 ) between states |# and |L does not generally commute with the axialvector coupling matrix: [X5a (6); M 2 ]#L = 0 :

(6.38)

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Then, the algebraic representations of the axialvector coupling matrix do not always coincide with the mass eigenstates: There occur representation mixings. Let us 0rst consider the zero helicity (6 =0) states and saturate the algebraic relation in Eq. (6.33) by low lying mesons; the , the (longitudinal) , the (longitudinal) axialvector meson denoted by A1 (a1 meson and its >avor partners) and the scalar meson denoted by S, and so on. The and the longitudinal A1 are admixture of (8; 1) ⊕ (1; 8) and (3; 3∗ ) ⊕ (3∗ ; 3), since the symmetry is spontaneously broken [87,186]: | = |(3; 3∗ ) ⊕ (3∗ ; 3) sin

+ |(8; 1) ⊕ (1; 8) cos

|A1 (6 = 0) = |(3; 3∗ ) ⊕ (3∗ ; 3) cos

;

− |(8; 1) ⊕ (1; 8) sin

;

(6.39)

where the experimental value of the mixing angle is given by approximately = =4 [87,186]. On the other hand, the longitudinal belongs to pure (8; 1) ⊕ (1; 8) and the scalar meson to pure (3; 3∗ ) ⊕ (3∗ ; 3): |(6 = 0) = |(8; 1) ⊕ (1; 8) ; |S = |(3; 3∗ ) ⊕ (3∗ ; 3) :

(6.40)

When the chiral symmetry is restored at the phase transition point, the axialvector coupling matrix commutes with the Hamiltonian matrix, and thus the chiral representations coincide with the mass eigenstates: The representation mixing is dissolved. From Eq. (6.39) we can easily see [106] that there are two ways to express the representations in the Wigner phase of the chiral symmetry: The conventional GL manifestation corresponds to the limit → =2 in which is in the representation of pure (3; 3∗ ) ⊕ (3∗ ; 3) [(Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf ) of SU(Nf )L × SU(Nf )R in large Nf QCD] together with the scalar meson, both being the chiral partners: (GL)

| ; |S → |(Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf ) ; |(6 = 0) ; |A1 (6 = 0) → |(Nf2 − 1; 1) ⊕ (1; Nf2 − 1) :

(6.41)

On the other hand, the VM corresponds to the limit → 0 in which the A1 goes to a pure (3; 3∗ ) ⊕ (3∗ ; 3) [(Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf )], now degenerate with the scalar meson in the same representation, but not with in (8; 1) ⊕ (1; 8) [(Nf2 − 1; 1) ⊕ (1; Nf2 − 1)]: (VM)

| ; |(6 = 0) → |(Nf2 − 1; 1) ⊕ (1; Nf2 − 1) ; |A1 (6 = 0) ; |S → |(Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf ) :

(6.42)

Namely, the degenerate massless and (longitudinal) at the phase transition point are the chiral partners in the representation of (8; 1) ⊕ (1; 8) [(Nf2 − 1; 1) ⊕ (1; Nf2 − 1)]. 57 Next, we consider the helicity 6 = ±1. As we stressed above, the transverse can belong to the representation di/erent from the one for the longitudinal (6 = 0) and thus can have the di/erent chiral partners. According to the analysis in Ref. [87], the transverse components of (6 = ±1) in the broken phase belong to almost pure (3∗ ; 3) (6 = +1) and (3; 3∗ ) (6 = −1) with tiny mixing with 57

We again stress that the VM is realized only as a limit approaching the critical point from the broken phase but not exactly on the critical point where the light spectrum including the and the would disappear altogether.

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(8; 1) ⊕ (1; 8). Then, it is natural to consider in VM that they become pure (Nf ; Nf∗ ) and (Nf∗ ; Nf ) in the limit approaching the chiral restoration point: |(6 = +1) → |(Nf∗ ; Nf ) ;

|(6 = −1) → |(Nf ; Nf∗ ) :

(6.43)

As a result, the chiral partners of the transverse components of in the VM will be themselves. Near the critical point the longitudinal becomes almost , namely the would-be NG boson almost becomes a true NG boson and hence a di/erent particle than the transverse . The A1 in the VM is resolved and/or decoupled from the axialvector current near the critical Davor [106] since there is no contribution in the vector current correlator to be matched with the axialvector current correlator. As to the scalar meson [97,98,115,124,149,181], although the mass is smaller than the matching scale adopted in Ref. [105] for Nf = 3, 58 we expect that the scalar meson is also resolved and/or decoupled near the chiral phase transition point [106], since it is in the (Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf ) representation together with the A1 in the VM. We further show the di/erence between the VM and GL manifestation discussed above in the quark contents. In the chiral broken phase, the pion and the axialvector meson couple to both the pseudoscalar density (q S 5 q) and the axialvector current (q S $ 5 q). On the other hand, the scalar meson couples to the scalar density (qq), S and the vector meson couples to the vector current (q S $ q). This situation is schematically expressed as q S 5 q ∼ G ⊕ GA A$ ; qq S ∼ GS S ; q S $ q ∼ FV V$ ; q S $ 5 q ∼ F ⊕ FA A$ :

(6.44)

In the GL manifestation, F becomes small and GS becomes identical to G near the restoration point. Then the scalar meson is a chiral partner of the pion. On the other hand, in the VM G becomes small and FV becomes identical to F . Thus the vector meson becomes a chiral partner of the pseudoscalar meson. The problem is which manifestation the QCD would choose. As we discussed in Section 6.1, the Wilsonian matching persists F () = 0, even at the critical point where F (0) = 0. Thus we conclude [106] that the VM is preferred by the QCD chiral restoration. 6.1.3. Conformal phase transition In this section, we shall argue that there actually exists an example of non-GL manifestation in 0eld theoretical models, which is called “conformal phase transition” [148] characterized by an essential-singularity-type scaling (see below). Following Ref. [148], we here brie>y summarize the “conformal phase transition”, and demonstrate how the GL (linear sigma model-like) manifestation breaks down, using the Gross–Neveu model [90] as an example. 58

The scalar meson does not couple to the axialvector and vector currents, anyway.

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In the linear sigma model-like phase transition, around the critical point z = zc (where z is a generic notation for parameters of a theory, as the coupling constant #, number of particle >avors Nf , etc), an order parameter P takes the form P = f(z)

(6.45)

( is an ultraviolet cuto/), where f(z) has a non-essential singularity at z = zc such that lim f(z) = 0 as z goes to zc both in the symmetric and broken phases. The standard form for f(z) is f(z) ∼ (z − zc ) , ¿ 0, around z = zc . The “conformal phase transition” is a very di/erent continuous phase transition. We de0ne it as a phase transition in which an order parameter P is given by Eq. (6.45) where however f(z) has an essential singularity at z = zc in such a way that while lim f(z) = 0

z →zc

(6.46)

as z goes to zc from the side of the broken phase, lim f(z) = 0 as z → zc from the side of the symmetric phase (where P ≡ 0). Notice that since relation (6.46) ensures that the order parameter P → 0 as z → zc , the phase transition is continuous. A typical example of the conformal phase transition is given by the phase transition in the (1 + 1)-dimensional Gross–Neveu model. Here we 0rst consider the dynamics in the D-dimensional (2 6 D ¡ 4) Nambu–Jona–Lasinio (Gross–Neveu) model, and then, describe the “conformal phase transition” in the Gross–Neveu (GN) model at D = 2. This will allow to illustrate main features of the “conformal phase transition” in a very clear way. The Lagrangian of the D-dimensional GN model, with the U (1)L × U (1)R chiral symmetry, takes the same form as Lagrangian (4.86) for the Nambu–Jona–Lasinio model in 4 dimensions: G L = S i $ 9$ + [( S )2 + ( S i 5 )2 ] ; (6.47) 2 where $ = 0; 1; : : : ; D − 1, and the fermion 0eld carries an additional “color” index # = 1; 2; : : : ; Nc . As we have shown in Eq. (4.87), the theory is equivalent to the theory with the Lagrangian 1 (’2 + 2 ) : L = S i $ 9$ − S (’ + i 5 ) − (6.48) 2G Let us look in this theory, which takes the same form as Eq. (4.88) % 4at the e/ective potential % D except that d k is replaced by d k. It is explicitly calculated as [133]: 2

4 1 1 4Nc D 2 6D D − +O ; (6.49) V (’; ) = + (4)D=2 5(D=2) g gcr 22 4 − D D 4 where = (’2 + 2 )1=2 , 6D = B(D=2 − 1; 3 − D=2), the dimensionless coupling constant g is de0ned by g=

4Nc D−2 G ; (4)D=2 5(D=2)

and the critical coupling gcr = D=2 − 1. At D ¿ 2, one 0nds that gcr − g d 2 V 4Nc D−2 2 : M’ ≡ d2 (4)D=2 5(D=2) gcr g =0

(6.50)

(6.51)

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As shown for the 4-dimensional NJL model in Eq. (4.105), the sign of M’2 de0nes two di/erent phases: M’2 ¿ 0 (g ¡ gcr ) corresponds to the symmetric phase and M’2 ¡ 0(g ¿ gcr ) corresponds to the broken phase with spontaneous chiral symmetry breaking, U (1)L ×U (1)R → U (1)L+R . The value M’2 = 0 de0nes the critical point g = gcr . Therefore at D ¿ 2, a linear-sigma-model-like phase transition is realized. However the case D = 2 is special: now gcr → 0 and 6D → ∞ as D → 2. In this case the e/ective potential is the well-known potential of the Gross–Neveu model [90]:

2 N c 2 N c 2 ln 2 + 1 : V (’; ) = (6.52) − 2g 2 The parameter M’2 is now d 2 V 2 M’ = → −∞ : d2

(6.53)

=0

Therefore, in this model, one cannot use M’2 as a parameter governing the continuous phase transition at g = gcr = 0: the phase transition is not a linear sigma model-like phase transition in this case. Indeed, as follows from Eq. (6.52), the order parameter, which is a solution to the gap equation dV=d = 0, is 1 S = exp − (6.54) 2g in this model. The function f(z), de0ned in Eq. (6.45), is now f(g) = exp(−1=2g), i.e., z = g, and therefore the conformal phase transition takes place in this model at g = 0: f(g) goes to zero only if g → 0 from the side of the broken phase (g ¿ 0). Let us discuss this point in more detail. At D ¿ 2, the spectrum of the ’ and excitations in the symmetric solution, with S = 0, is de0ned by the following equation (in leading order in 1=Nc ) [133]: 1 1 6D − (−M2 )D=2−1 = 0 : D−2 + (6.55) g gcr 2 − D=2 Therefore at D ¿ 2, there are tachyons with 4 − D 2=(D−2) g − gcr 2=(D−2) 2 2 2 2 M = M’ = Mtch = − 26D gcr g at g ¿ gcr , and at g ¡ gcr there are “resonances” with 4 − D 2=(D−2) gcr − g 2=(D−2) 2 2 2 : |M | = |M’ | = 26D gcr g

(6.56)

(6.57)

Eq. (6.57) implies that the limit D → 2 is special. One 0nds from Eq. (6.55) that at D = 2 1 2 2 2 2 (6.58) M = M’ = Mtch = − exp − g at g ¿ 0, and |M2 |

=

|M’2 |

1 = exp |g| 2

(6.59)

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at g ¡ 0, i.e., in agreement with the main feature of the conformal phase transition, there are no light resonances in the symmetric phase at D = 2. The e/ective potential (6.52) can be rewritten as

2 Nc 2 ln 2 − 1 (6.60) V (’; ) = 2 S (with S given by Eq. (6.54)) in the broken phase. That is, in this phase V (’; ) is 0nite in the continuum limit → ∞ after the renormalization of the coupling constant, 1 (6.61) g= ln (2 = S2 ) [see Eq. (6.54)]. But what is the form of the e/ective potential in the continuum limit in the symmetric phase, with g ¡ 0? As Eq. (6.52) implies, it is in0nite as → ∞: indeed at g ¡ 0, there is no way to cancel the logarithmic divergence in V . It is unlike the case with D ¿ 2: in that case, using Eq. (6.51), the potential (6.49) can be put in a linear-sigma-model-like form: M’2 2 6D 8Nc V (’; ) = + D : (6.62) 2 (4)D=2 5(D=2) (4 − D)D However, since M’2 =−∞ at D=2, the linear-sigma-model-like form for the potential is not available in the Gross–Neveu model. What are physical reasons of such a peculiar behavior of the e/ective potential at D = 2? Unlike the case with D ¿ 2, at D=2, Lagrangian (6.47) de0nes a conformal theory in the classical limit. By using the conventional approach, one can derive the following equation for the conformal anomaly in this model (see, for detailed derivation, the appendix of Ref. [148]): 9$ D$ = :$$ = L(g)[( S )2 + ( S i 5 )2 ] ; (6.63) 2Nc

where D$ is the dilatation current, :$ is the energy–momentum tensor, and the L(g) is the L function given by 9g = −g2 (6.64) L(g) = 9 ln both in the broken and symmetric phases. While the broken phase (g ¿ 0) corresponds to asymptotically free dynamics, the symmetric phase (g ¡ 0) de0nes infrared free dynamics: as → ∞, we are led to a free theory of massless fermions, which is of course conformal invariant. On the other hand, in the broken phase the conformal symmetry is broken, even as → ∞. In particular, Eq. (6.60) implies that 2Nc 2 S = 0 0|:$$ |0 = 4V () S =− (6.65) in leading order in 1=Nc in that phase. The physics underlying this di/erence between the two phases in this model is clear: while g ¡ 0 corresponds to repulsive interactions between fermions, attractive interactions at g ¿ 0 lead to the formation of bound states, thus breaking the conformal symmetry. Thus the conformal phase transition describes the two essentially di/erent realizations of the conformal symmetry in the symmetric and broken phases.

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The conformal phase transition is also observed in other 0eld theoretic models: A most notable example is the ordinary QCD (with small Nf ) which exhibits a well-known essential-singularity-type scaling at #() = 0: m ∼ e−1=b#() ;

(6.66)

although it has no symmetric phase (corresponding to # ¡ 0). Similar essential-singularity-type scaling has been observed in the ladder QED [145], the gauged NJL model in the ladder approximation [132,13], etc. We shall discuss in Section 6.2 a conformal phase transition observed in the large Nf QCD within the ladder approximation. (Details are discussed in Ref. [148].) 6.1.4. Vector manifestation vs. “vector realization” The VM in the HLS is similar to the “vector realization” [85,86] also formulated in the HLS, in the sense that the chiral symmetry gets unbroken in such a way that vector meson becomes massless m → 0 and a chiral partner of . However VM is di/erent from the “vector realization” in an essential way: The “vector realization” was claimed to be neither the Wigner realization nor the NG realization in such a way that the NG boson does exist (F (0) = 0), while the chiral symmetry is still unbroken (qq S = 0): F (0) = 0;

qq S =0 :

(6.67)

On the contrary, our VM is precisely the limit of the Wigner realization having F (0) = 0;

qq S =0 :

(6.68)

A crucial di/erence between the two comes from the fact that in VM the quadratic divergence of our Wilsonian RGEs leads to the Wigner realization with F (0) → 0 at the low-energy limit (on-shell of NG bosons) in spite of F () = 0, while in the “vector realization” the quadratic divergence is not included and hence it was presumed that F (0) = F () and thus F (0) = 0. Technically, in the vector limit (or the VM limit with the VM conditions), the bare HLS Lagrangian in the VM and that of the “vector realization”, formally approach the same 0xed point Lagrangian L∗HLS which is de0ned just on the 0xed point g() = 0, a() = 1 and F () = 0 (plus z1 () = z2 ()): ˆ$ V ˆ $ ] + tr[Aˆ$ Aˆ$ ]} L∗HLS = F2 () {tr[#ˆ⊥$ #ˆ$ ] + tr[#ˆ $ #ˆ$ ]} + z1 () {tr[V ⊥

z1 () F2 () ˆ $ L ˆ $ ] + [Rˆ $ Rˆ $ ]} ; (6.69) tr{[D$ CL · C†L ]2 + [D$ CR · C†R ]2 } + tr{[L 4 2 where D$ CL ≡ 9$ C + iCL$ (and L ↔ R). However, when the external gauge 0elds are switched o/, it was pointed out [85,86] that the 0xed point Lagrangian L∗HLS possesses a large (global) symmetry based on the manifold G1 × G2 [SU(Nf )L × SU(Nf )R ]1 × [SU(Nf )L × SU(Nf )R ]2 ; (6.70) = G SU(Nf )L1 +L2 × SU(Nf )R1 +R2 =−

where the residual symmetry G = SU(Nf )L1 +L2 × SU(Nf )R1 +R2 was identi0ed in Ref. [85,86] with the chiral symmetry of the QCD, while G1 × G2 is a (global) symmetry larger than that of QCD such that CL → gL1 CL gL† 2 ;

(6.71)

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with gL1 ∈ SU(Nf )L1 and gL2 ∈ SU(Nf )L2 (and L ↔ R). Then the 0xed point Lagrangian L∗HLS has no connection with the QCD and must be decoupled from QCD! Even if we are o/ the point (a(); g()) = (1; 0) by a() = 1, we still have a redundant global symmetry H × G which is larger than the QCD symmetry by the additional global symmetry H (⊂ G1 ), where G1 is reduced to the subgroup H by a() = 1. When the HLS coupling is switched on, g() = 0, on the other hand, the G1 (or H ⊂ G1 when a() = 1) becomes a local symmetry, namely the HLS Hlocal = SU(Nf )local , and hence the larger global symmetry G1 × G2 is reduced to the original symmetry of the HLS model, Hlocal × Gglobal (G=G2 ), as it should, in accord with the QCD symmetry. Such a redundant larger (global) symmetry G1 × G2 (or H × G) is speci0c to just on the 0xed point g ≡ 0; a ≡ 1 (or g ≡ 0). Then the point (a; g) = (1; 0) must be regarded only as a limit g( = 0) → 0 ;

(6.72)

in which case the e/ective Lagrangian has no such a redundant global symmetry. Actually, as was shown in Section 5.3.2, the real-life QCD with Nf = 3 is very close to a() = 1 but g2 ()1, which means that Nature breaks such a redundant G1 × G2 symmetry only by a strong coupling gauge interaction of the composite gauge boson . When we approach the chiral restoration point of the underlying QCD, this strong gauge coupling becomes vanishingly small, thus forming a weak coupling composite gauge theory, but the gauge coupling should never vanish, however small. In the next sub-subsection, we shall discuss in detail that VM must actually be regarded as such a limit. On the other hand, situation is completely di/erent for the “vector realization”: In order to have the unbroken chiral symmetry of QCD under the condition F (0) = 0, namely existence of NG bosons, and , it desperately needs a redundant larger global symmetry which is to be spontaneously broken down to the unbroken chiral symmetry of QCD. It then must be formulated precisely on the point (a; g) ≡ (1; 0) whose e/ective Lagrangian L∗HLS in Eq. (6.69) actually does have such a redundant symmetry. Then it implies that “vector realization” is decoupled from the QCD! We now show, based on the general arguments [200] on the chiral Ward–Takahashi (WT) identity, that the “vector realization”, Eq. (6.67), implies that the NG bosons are actually all decoupled from the QCD. This is consistent with the fact that the 0xed point Lagrangian in the “vector realization” has a di/erent symmetry than QCD and is decoupled from the QCD. Let us start with the symmetry G of a system including 0elds =i under the transformation A =i = −i(T A )ij =j = [iQA ; =i ], with A = 1; 2; : : : ; dim G, where T A are the matrix representations of the generators of the symmetry group G and QA the corresponding charge operators. Let the symmetry be spontaneously broken into a subgroup H , Qa |0 = 0, where Qa are the charges corresponding to the broken generators T a ∈ G − H, with G and H being algebras of G and H , in such a way that a Gn (x1 ; : : : ; x n ) = 0|[iQa ; T =1 (x1 ) · · · =n (x n )]|0 = 0 ;

(6.73)

where a Gn is an n-point order parameter given by the variation of the n-point Green function Gn (x1 ; : : : ; x n ) ≡ 0|T=1 (x1 ) · · · =n (x n )|0

(6.74)

(T : time-ordered product) under the transformation corresponding to the broken generators T a . Then, we have a general form of the chiral WT identity: lim q$ M$a = a Gn (x1 ; : : : ; x n ) ;

q$ →0

(6.75)

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where the current-inserted Green function for broken current (axialvector current) J5$ is de0ned by

a a (z)=1 (x1 ) · · · =n (x n )|0 : (6.76) M$ (q; x1 ; : : : ; x n ) ≡ d 4 z eiqz 0|T J5$ Noticing that a Gn (x1 ; : : : ; x n ) is a residue of the NG boson pole at q2 = 0 in M$a (q; x1 ; : : : ; x n ), we have [200] a Gn (x1 ; : : : ; x n ) = F (0) · a (q$ = 0)|T =1 (x1 ) · · · =n (x n )|0 ;

(6.77)

where a (q$ )|T =1 (x1 ) · · · =n (x n )|0 is a Bethe–Salpeter amplitude which plays a role of “wave function” of the NG boson a and the NG boson decay constant F (0) is de0ned by 0|J$a (x)|b (q) = −iab F (0)q$ e−iqx :

(6.78)

A simple example of the relation Eq. (6.77) is given by the linear sigma model: a b = ab and =−a , a G1 (x)=a 0|b (x)|0 =ab 0|(x)|0 , while a |b (x)|0 =ab Z1=2 , and hence Eq. (6.77) reads 0||0 = F (0) Z1=2 , or = F (0) at tree level where the wave function renormalization is trivial, Z1=2 = 1. Another popular example is the (generalized) Goldberger–Treiman relation for the quark propagator S(p) = FTG2 (x) = FT0|Tq(x)q(0)| S where FT stands for the Fourier transform: Eq. (6.77) reads a G2 (x) = F (0) · a (q$ = 0)|Tq(x)q(0)|0 , S which after taking Fourier transform reads 59 a

2 O(p2 ) = F (0) 5a (p; 0) ;

(6.79)

where O(p2 ) is the dynamical mass of the quark parametrized as iS −1 (p) = Z −1 ( $ p$ − O(p2 )) and a S =S {−i 5 T a ; S −1 } S = 5 T a Z −1 S ·2 O ·S for a q(x)=−i 5 T a q(x), while 5a (p; q) is an amputated (renormalized) Bethe–Salpeter amplitude of a , or dynamically induced –q–q vertex; 5 T a Z −1 S(p + q) 5a (p; q) S(p) ≡ FTa (q$ )|Tq(x)q(0)|0 S :

(6.80)

Now, when the broken symmetry is restored, Qa |0 = 0, we simply have a Gn (x1 ; : : : ; x n ) = 0|[iQa ; T =1 (x1 ) · · · =n (x n )]|0 = 0

(6.81)

for all Green functions. If one assumed there still exist NG bosons F (0) = 0 as in “vector realization”, then Eq. (6.77) would dictate F (0) · a (q$ = 0)|T =1 (x1 ) · · · =n (x n )|0 = 0 ;

(6.82)

and hence a (q$ = 0)|T =1 (x1 ) · · · =n (x n )|0 = 0

for all n :

(6.83)

This would imply a situation that the NG bosons with q$ = 0 would be totally decoupled from any operator, local or nonlocal, of the underlying theory, the QCD in the case at hand. Then the “vector realization” is totally decoupled from the QCD, which is also consistent with the fact that the 0xed point Lagrangian Eq. (6.69) has a di/erent symmetry than QCD. 59

In the case of -nucleon system, 5a (p; 0) reads GNN (NN Yukawa coupling) and O(p2 ) does mN (nucleon mass), and hence the Goldberger–Treiman relation follows 2mN gA = F (0) GNN with gA = 1. gA = 1 would follow only when we take account of the fact that the nucleon is not the irreducible representation of the chiral algebra due to the representation mixing in the Adler–Weisberger sum rule.

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On the other hand, the VM is simply a limit to a Wigner phase, F (0) → 0 ;

(6.84)

and hence we can have a (q$ = 0)|T =1 (x1 ) · · · =n (x n )|0 = 0 although a in this case are no longer the NG bosons and may be no longer light composite spectrum as in the conformal phase transition [148]. If the light composite spectrum disappear as in conformal phase transition, then the e/ective 0eld theory breaks down anyway just at the phase transition point. 6.1.5. Vector manifestation only as a limit We actually de0ned the VM as a limit (“VM limit”) with bare parameters approaching the 0xed point, VM point (X (); a(); G()) = (1; 1; 0) = (X2∗ ; a∗2 ; G2∗ ), from the broken phase but not exactly on the 0xed point. Since the 0xed point Lagrangian has a di/erent symmetry than QCD, we must approach the VM limit along the line other than G ≡ 0 (Fig. 12 in Section 4.10). We shall give an example to approach the VM limit from G = 0 in Section 6.3.2. Here we demonstrate through the chiral WT identity that a relation precisely on the point g = 0 contradicts the QCD even when F2 (0) → 0, while that as a limit g → 0 is perfectly consistent. This also gives another example to show that the “vector realization” is decoupled from the QCD. The chiral WT identity is the same as that in the previous sub-subsection except that two axialvector currents J5$ and two vector currents J$ are involved: ab;cd c d (x1 ; x2 ; q1 ; q2 ) = FT0|TJ5$ (z1 )J5 (z2 )J#a (x1 )JLb (x2 )|0 ; M#L;$

(6.85)

where FT stands for Fourier transform with respect to z1 and z2 . Then we have lim

q1 →0; q2 →0

ab;cd q1$ q2 M#L;$ (x1 ; x2 ; q1 ; q2 )

=(face fbde + fade fbce )(0|TJ5# (x1 )J5L (x2 )|0 − 0|TJ# (x1 )JL (x2 )|0 ) ;

(6.86)

where use has been made of a b 0|TJ5# (x1 )J5L (x2 )|0 = ab 0|TJ5# (x1 )J5L (x2 )|0 ;

(6.87)

ab;cd , we have etc. Looking at the residues of massless poles of two ’s in M#L;$ ab;cd (x1 ; x2 ; 0; 0) F2 (0) · 5#L

=(face fbde + fade fbce )(0|TJ5# (x1 )J5L (x2 )|0 − 0|TJ# (x1 )JL (x2 )|0 ) ;

(6.88)

where ab;cd 5#L (x1 ; x2 ; q1 ; q2 ) = c (q1 )|TJ#a (x1 )JLb (x2 )|d (q2 )

(6.89)

is the amplitude for process. Taking Fourier transform with respect to x1 − x2 and omitting a, b, c and d, we have F2 (0) · 5˜ #L (k) = (g#L k 2 − k# kL )[TA (k 2 ) − TV (k 2 )] :

(6.90)

˜ 2 ), we have Writing 5˜ #L (k) = (g#L − k# kL =k 2 )5(k ˜ 2 ) = k 2 [TA (k 2 ) − TV (k 2 )] ; F2 (0) · 5(k

(6.91)

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which clearly shows that when the chiral symmetry gets restored as TA (k 2 ) − TV (k 2 ) → 0 in the ˜ 2 ) → 0 for any k 2 , i/ F (0) = 0 as in the “vector underlying QCD, we would have a disaster, 5(k 2 realization”. Actually, by taking a limit k → 0, we have ˜ = F2 (0) ; F2 (0) · 5(0)

(6.92)

where use has been made of the Weinberg 0rst sum rules, limk 2 →0 k 2 [TA (k 2 ) − TV (k 2 )] = F2 (0), which are valid in QCD for Nf ¡ 33=2 even in the restoration limit F (0) → 0. Then it follows ˜ 5(0) =1 ;

(6.93)

as far as F (0) = 0 (including the limit F (0) → 0). ˜ 2 ) in terms of the HLS model at O(p2 ): Now we compute the amplitude 5(k ˜ 2 ) = (1 − a) + a 5(k

M2 ; M2 − k 2

(6.94)

the 0rst term of which corresponds to the direct coupling of , while the second term does to the vertex followed by the transition → . (There is no vertex in the HLS model at leading order.) If we set g ≡ 0 (hence M2 ≡ 0), then we would get ˜ 2) = 1 − a 5(k

(6.95)

for all k 2 , which would vanish at a → 1 in contradiction with the QCD result, Eq. (6.93). This again implies that the “vector realization” with F2 (0) = 0 is inconsistent with QCD. Eq. (6.95) implies that the VM having F2 (0) → 0 is also inconsistent with Eq. (6.93), although not inconsistent with the Eq. (6.92). On the other hand, if we take a limit g → 0 at k 2 = 0 for the VM, we have ˜ 5(0) =1 ;

(6.96)

in perfect agreement with Eq. (6.93). Therefore the VM must be formulated as a limit g() ( = 0) → 0

(6.97)

in such a way that F2 (0) ( = 0) → 0, while we can safely put a = 1. Similar unphysical situation can be seen for the parameter X de0ned in Section 4.10: When the bare parameter X () approaches the one at the VM point (X2∗ ; a∗2 ; g2∗ ) = (1; 1; 0) from the broken phase as X (), g() → 0, the parameter X (0) approaches 0 as X (0) → 0, which implies that m2 =F2 (0) → 0 [see Section 6.3.2]. When the theory is exactly on the VM point, on the other hand, we have X () ≡ 1 which leads to X (0) = 1 since (X2∗ ; a∗2 ; g2∗ ) = (1; 1; 0) is the 0xed point. The discussion in this subsection also implies that presence of gauge coupling, however small, can change drastically the pattern of symmetry restoration in the nonlinear sigma model: For instance, the lattice calculation has shown that the Nf = 2 chiral Lagrangian has a O(4) type restoration, i.e., the linear sigma model-type restoration, while it has not given a de0nite answer if it is coupled to gauge bosons like , namely the lattice calculation has not been inconsistent with the VM, other than O(4)-type restoration, in the limit g → 0 (not g ≡ 0) even for Nf = 2. 60 60

We thank Yoshio Kikukawa for discussions on this point.

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6.2. Chiral phase transition in large Nf QCD In this subsection we summarize the known results of the chiral symmetry restoration in the large Nf QCD, with Nf (Nf ¡ Nf∗∗ ≡ 11 N ) being number of massless quark >avors. For a certain large 2 c Nf the coupling has an infrared 0xed point which becomes very small near 11 N [26]. The two-loop 2 c L function is given by L(#) = −b#2 − c#3 ;

(6.98)

where two coeVcients are [125,51]: 1 (11Nc − 2Nf ) ; b= 6 Nc2 − 1 1 2 34Nc − 10Nc Nf − 3 (6.99) Nf : c= 242 Nc There is at least one renormalization scheme in which the two-loop L function is (perturbatively) exact [179]. We will use such a renormalization scheme. Then we have an infrared 0xed point for if b ¿ 0 and c ¡ 0 b (6.100) # = #∗ = − : c When Nf is close to (but smaller than) Nf∗∗ = 11 N , the value of #∗ is small and hence one should 2 c expect that the chiral symmetry is not spontaneously broken: namely, there is a critical value of Nf , Nf = Nfcrit beyond which the (spontaneous broken) chiral symmetry is restored [14]. Actually when we decrease Nf , the value of the 0xed point #∗ increases and eventually blows up (this 0xed point disappears) at the value Nf =Nf∗ when the coeVcient c becomes positive (Nf∗ 8:05 for Nc = 3, although this value is not reliable since the perturbation must breaks down for strong coupling). However, before reaching Nf∗ the perturbative infrared 0xed point in the L function will disappear at Nf = Nfcrit (¿ N ∗ ) where the coupling #∗ exceeds a certain critical value #c so that the chiral symmetry is spontaneously broken; namely, fermions can acquire a dynamical mass and hence decouple from the infrared dynamics, and only gluons will contribute to the L function. The value Nfcrit may be estimated in the (improved) ladder Schwinger–Dyson (SD) equation combined with the perturbative 0xed point [14]. It is well known [139,75,73] that in the (improved) ladder SD equation the spontaneous chiral symmetry breaking would not occur when the gauge coupling is less than a critical value # ¡ #∗ ¡ #c = (2Nc =(Nc2 − 1)) · =3. Then, the estimate for the critical value Nfcrit is given by [14]: or,

#∗ |Nf =Nfcrit = #c Nfcrit

= Nc

100Nc2 − 66 25Nc2 − 15

(6.101) 12

Nc : 3

(6.102)

However, the above estimate of Nfcrit through the ladder SD equation combined with the perturbative 0xed point may not be reliable, since besides various uncertainties of the ladder approximation for the estimate of the critical coupling #c , the perturbative estimate of the 0xed point value #∗ in Eq. (6.100) is far from reliable, when it is equated to #c which is of order O(1).

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As we said before, such a chiral symmetry restoration in the large Nf QCD is actually observed by various other methods such as the lattice simulation [131,41,61,117–122], dispersion relation [153,154], instanton calculus [182], etc. The most recent result of the lattice simulation shows [118] 6 ¡ Nfcrit ¡ 7 ;

(6.103)

which is substantially smaller than the ladder-perturbative estimate Eq. (6.102). Although the ladder-perturbative estimate of Nfcrit may not be reliable, it is worth mentioning that the result of the ladder SD equation has a scaling of an essential-singularity for the dynamical mass m of the fermions [14], called “Miransky scaling” as 0rst observed in the ladder QED [145]:   − −C  ; m ≈ exp (6.104) = exp #∗ =#c − 1 1=Nf − 1=Nfcrit with being the “cuto/” of the dominant momentum region in the integral of the SD equation and C = (13Nc2 Nf − 34Nc3 − 3Nf )=(100Nc3 Nf − 66Nc Nf ). Relatively independent of the estimate of Nfcrit , this feature may describe partly the reality of the chiral phase transition of large Nf QCD. It was further shown [14,148] in the ladder approximation that the light spectrum does not exist in the symmetric phase in contrast to the broken phase where the scalar bound state becomes massless in addition to the massless NG boson . As was discussed in Section 6.1.3, these features are in accord with the conformal phase transition where the Ginzburg–Landau (GL) e/ective theory (linear sigma model-like manifestation, or GL manifestation) simply breaks down. It is also to be noted that the (two-loop) running coupling in this theory is expected to become walking, #(Q2 ) #∗ for entire low energy region Q2 ¡ 2 , so that the condensate scales with anomalous dimension m 1 as in the walking technicolor [113,202,4,11,25] (for reviews see Refs. [200,112]): qq S ∼ m2 ;

(6.105)

with m given by Eq. (6.104). 6.3. Chiral restoration and VM in the eAective ?eld theory of large Nf QCD In this subsection we show that the chiral restoration in the large Nf QCD, F2 (0) → 0, is also derived in the EFT, the HLS model, when we impose the Wilsonian matching to determine the bare parameters by the VM conditions, Eqs. (6.11)–(6.14). Once the chiral restoration takes place under the VM conditions, the VM actually occurs at the critical point as we demonstrated in Section 6.1.1 and so does the VM in the large Nf QCD. It is to be noted that although the HLS model as it stands carries only the information of Nf of the underlying QCD but no other information such as Nc and QCD , the latter information actually is mediated into the bare parameters of the HLS model, F2 (), g(), a(), etc., through the Wilsonian matching. Then we can play with Nc and QCD as well as Nf even at the EFT level. 6.3.1. Chiral restoration As we have already shown in Section 6.1.1, when the chiral restoration takes place in the underlying QCD, we have the VM conditions which lead to Eq. (6.15): N f 2 F2 (0) → F2 () − ; (6.106) 2(4)2

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with F2 () being given by Eq. (6.14): Nc 2 2 crit 2 · 2(1 + crit F () → (F ) = A ) : 3 4

155

(6.107)

Then the chiral restoration can take place also in the HLS model when F2 (0) = (Fcrit )2 − or X () ≡

Nf 2 →0 ; 2(4)2

(6.108)

Nf 2 N f 2 1 1 → →1 : 2 2 2 crit 2(4) F () 2(4) (F )2

(6.109)

This is actually realized in a concrete manner in the HLS model for the large Nf QCD. In the large Nf QCD at the chiral restoration point, F2 () determined by the underlying QCD is almost independent of Nf but crucially depends on (and is proportional to) Nc , while the quadratic divergence of the HLS model does on Nf . Then Nf is essentially the only explicit parameter of the HLS model to be adjustable after VM conditions Eqs. (6.12)–(6.14) are imposed and can be increased for 0xed Nc towards the critical Nf : Nf → Nfcrit − 0 ;

(6.110)

(Fcrit )2 Nc = (6.111) · 4(1 + crit A ) : 2 3 Note that this corresponds to X () → 1 − 0 in accord with the >ow in Fig. 13: If we take X () → 1+0, on the other hand, we would enter, before reaching the VM 0xed point (1; 1; 0), the symmetric phase where the HLS model breaks down or the light composite spectrum would disappear in the underlying QCD with Nf ¿ Nfcrit . As will be discussed below, Nfcrit = 2(4)2

crit S =0 = A ≡ A |qq

3(Nc2 − 1) #s 22 (#s =)G$ G $ + 8Nc Nc 4f

(6.112)

is almost independent of Nc as well as Nf , and is roughly given by simply neglecting the quark condensate term (the third term) in A |Nc =Nf =3 ( 0:36) in Eq. (5.20): crit A 0:27 ± 0:04 ± 0:03 ; which yields Nfcrit

(5:1 ± 0:2 ± 0:1) ×

(6.113)

Nc 3

;

(6.114)

where the center values in Eqs. (6.113) and (6.114) are given for (3 ; QCD ) = (1:1; 0:4) GeV, and the 0rst and second errors are obtained by allowing 3 and QCD to vary 3 = 0:1 GeV and QCD = 0:05 GeV, respectively. Hence Eq. (6.111) implies that Nfcrit ∼ O(Nc ). This is natural, since both F2 (0) and F2 () are of O(Nc ) in Eq. (6.106) (see the discussion in Section 5.1) and so is the Nf as far as it is to be non-negligible (near the critical point). Thus the chiral restoration is a peculiar phenomenon which takes place only when both Nf and Nc are regarded as large, with Nf ∼ Nc 1 :

(6.115)

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Historically, the chiral restoration in terms of HLS for the large Nf QCD was 0rst obtained in Ref. [104], based on an assumption that the bare parameters take the 0xed point values (a; g) = (1; 0) and hence Eq. (6.15) follows and also based on a further assumption that F ()2 =2 has a small dependence on Nf (for 0xed Nc ). These assumptions were justi0ed later by the Wilsonian matching [105,106], although the second assumption got a small correction between Nf = Nfcrit and Nf = 3 essentially arising from a factor (1 + crit A )=(1 + A |Nc =Nf =3 ) ∼ (1 + 0:27)=(1 + 0:36) ∼ 0:93. Now, we discuss the result of Nfcrit in Eq. (6.114) based on the estimation of Eq. (6.112). First of all we should mention that the OPE is valid only for crit A (¡ A ) ¡ 1 and hence in order for our approach to be self-consistent, our estimate of Nfcrit must be within the range: Nc Nc Nc crit crit ¡ Nf = ; (6.116) 4 · 4(1 + A ) ¡ 8 3 3 3 which is consistent with the recent lattice result 6 ¡ Nfcrit ¡ 7 (for Nc =3) [118] and in sharp contrast to the ladder-perturbative estimate Nfcrit 12Nc =3 [14]. Let us next discuss some details: Here we make explicit the Nf -dependence of the parameters like the matching scale f ≡ (Nf ) for 0xed Nc , #s (f ; Nf ), etc., since they generally depend on Nf (also on Nc and QCD from the QCD side). For the 0rst term in Eq. (6.112), as we will discuss later, Nc #s (f ; Nf ) is independent of Nf and Nc , and hence f increases with Nf and decreases with Nc . Then we use (Nc =3)#s (f ; Nf )==#s (3 ; Nf =3)=|Nc =3 0:22, again the value obtained in Eq. (5.20). For the gluonic condensate term, (#s =)G$ G $ is independent of the renormalization point of QCD, so that it is natural to say that it is independent of Nf . Furthermore, (1=Nc )(#s =) G$ G $ is independent of Nc [28]. Although f is somewhat larger than 3 as mentioned above, we here make a crude estimate of the second term by simply taking f =3 , (3 ; QCD )=(1:1; 0:40) GeV and using (3=Nc )(#s =)G$ G $ = 0:012 GeV4 [171,172,28], which yields the value, 0:054, already given for Nc = Nf = 3 in Eq. (5.20). At any rate, the gluon condensate term is numerically negligible (less than 5% for Nfcrit ) in any estimate and hence does not give much uncertainty. Now, we set (Nc2 −1)=Nc2 =8=9 but this factor will yield 1 for large Nc and thus enhance 0:22 to 0:25. In conclusion we have crit + 0:054 0:27(0:30), which yields A 0:22(0:25) N c (Nc ∼ 3) ; Nfcrit 5:1 3 Nc (Nc 3) : 5:2 (6.117) 3 A more precise estimation of Nfcrit will be done by determining the Nf -dependences of the QCD coupling #s and f in Section 6.3.3. Here we just quote the result Nfcrit 5:0Nc =3 (for Nc =3), which is consistent with the above estimate and somewhat similar to the recent lattice result 6 ¡ Nfcrit ¡ 7 (for Nc = 3) [118], while much smaller than the ladder-perturbative estimate Nfcrit 12Nc =3 [14]. It is amusing that our estimate coincides with the instanton argument [182]. If such a relatively small value of Nfcrit is indeed the case, it would imply that for some (nonperturbative?) reason the running coupling might level o/ in the infrared region at smaller Nf than that expected in the perturbation. At any rate, what we have shown here implies a rather amazing fact: Recall that the real-life QCD with Nf = 3 is very close to a() = 1 (see Section 5.3.2), which corresponds to the ideal situation

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157

that the bare HLS Lagrangian is the 0xed point Lagrangian, Eq. (6.70), having a redundant global symmetry G1 × G2 which is explicitly broken only by the strong gauge coupling. Now in the large Nf QCD with Nf very close to the critical point Nfcrit , the coupling becomes vanishingly small and the bare HLS Lagrangian realizes a weak coupling gauge theory with the G1 × G2 symmetry explicitly broken only by the “weak” coupling of the composite meson. 6.3.2. Critical behaviors In this section we study the critical behaviors of the parameters and several physical quantities when Nf approaches to its critical value Nfcrit using the RGEs. As we discussed at the end of Section 4.10, since the VM 0xed point (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0) is not an infrared stable 0xed point, the VM limit with bare parameters approaching the VM 0xed point from the broken phase does not generally imply that the parameters in the infrared region approach the same point: We expect that, without extra 0ne tuning, g2 (m ) → 0 is obtained from one of the VM conditions, g2 () → 0. Combining this with the on-shell condition (4.217) leads to the infrared parameter X (m ) behaving as X (m ) → 0, although X () → 1. This implies m2 =F2 (m ) → 0. From this together with Eq. (4.222) we infer m2 →0 : F2 (0)

(6.118)

Below we shall discuss that this is indeed the case by examining the critical behaviors of the physical parameters near the critical point in a more precise manner through the RGEs. For that we need to know how the bare parameters g(f ; Nf ) and a(f ; Nf ) approach to the VM limit in Eqs. (6.11) and (6.12). Taking the limits g2 ()1, M2 ()=2 = g2 ()a()F2 ()=2 1 and F2 ()=F2 () − 1 = a() − 11 in the Wilsonian matching condition (5.7), we obtain 2 2 F () F2 () F2 () 2 2 g () − (a() − 1) + 2g ()z () − 2[z2 () − z1 ()] 3 2 2 2 =

4(Nc2 − 1) #s qq S 2 : Nc2 6

(6.119)

It is plausible to require that there are no cancellations among the terms in the left-hand side (LHS) of the above matching condition. Then, we expect that all the terms in the LHS have the same scaling behavior near the restoration point. The critical behavior of the HLS gauge coupling g2 (f ; Nf ) is then given by #s g2 (f ; Nf ) ∼ 2 qq S 2 ; (6.120) Nc where we put the extra Nc -dependence coming from [F2 ()]2 ∼ Nc2 into the right-hand side of the above relation. Since the quark condensate scales as Nc , qq S ∼ Nc , and the QCD gauge coupling scales as 1=Nc in the large Nc counting, the above relation implies that the HLS gauge coupling scales as 1=Nc , g2 (f ; Nf ) ∼ 1=Nc , in the large Nc counting. Now we consider the Nf -dependence. We may parameterize the scaling behavior of g2 as 1 1 g2 (f ; Nf ) = gS 2 f(j) ; j ≡ − crit ; (6.121) Nf Nf

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where gS is independent of Nf . f(j) is a certain function characterizing the scaling of qq S 2 ∼ m6−2 m 2 m ;

(6.122)

where m is the anomalous dimension and m = m(j) is the dynamical mass of the fermion which vanishes as j → 0. For example, the improved ladder SD equation with the two-loop running gauge coupling [14] √ implies the walking gauge theory [113,202,4,11,25] which √ suggests m 1 and m = exp[ − C= j] as in Eq. (6.104), so that we have f(j) = exp[ − 4C= j]. However, since we do not know the reliable estimate of the scaling function f(j), we will leave it unspeci0ed in the below. To make an argument based on the analytic solution, we here 0x a(f ; Nf ) = 1

(6.123)

even o/ the critical point, since the Wilsonian matching conditions with the physical inputs F (0) = 86:4 MeV and m = 771:1 MeV leads to a() 1 already for Nf = 3 [see Section 5 as well as Ref. [105]]. Recall that putting a = 1 does not contradict the symmetry of the underlying QCD though g = 0 does (see Section 6.1.5). A deviation from a() = 1 will be discussed in the next sub-subsection. Before studying the critical behaviors of the parameters in the quantum theory, let us show the solutions of the RGEs (4.208) and (4.211). We note that these RGEs are solvable analytically when we take a = 1 from the beginning. From Eq. (4.211) with a = 1 the solution g2 ($; Nf ) is expressed as 1 g2 ($; Nf ) = ; (6.124) Cf b ln($=H (Nf )) where Cf = Nf =(2(4)2 ) and b = 43=3. H (Nf ), which generally depends on Nf , is the intrinsic scale of the HLS, analog to QCD of QCD. To show the solution for F ($; Nf ) it is convenient to use a cuto/ scale f as the reference scale. The solution is given by

l 2 l 2

s F2 ($; Nf ) F (f ; Nf ) g (f ; Nf ) Nf t (6.125) = − dz e−2z ; g2 ($; Nf ) (4)2 0 t − z 2f 2f where l = 9=43, s = ln(f =$) and t = ln(f =H (Nf )). Let us now study the critical behaviors of the parameters in the quantum theory. The solution (6.124) for g2 with Eq. (6.121) determines the critical behavior of the intrinsic scale of the HLS as Nf → Nfcrit : H (Nf ) → exp[ − T=f(j)], where T = 1=(Cf bgS 2 ). The intrinsic scale of the HLS goes to zero with an essential singularity scaling. Since m (Nf ) ¿ H (Nf ), it is natural to assume that the gauge coupling at the scale m (Nf ) approaches to zero showing the same power behavior: g2 (m (Nf ); Nf ) → gS 2 f(j) as Nf → Nfcrit . Replacing s with sV ≡ ln(f =m (Nf )) in Eq. (6.125) and substituting it into the on-shell condition (4.217), we obtain

l m2 (Nf ) g2 (f ; Nf ) 2 = g (m (Nf ); Nf ) 2 g (m (Nf ); Nf ) 2f ! l

sV Nfcrit Nf Nf t × − − dz − 1 e−2z 2(4)2 2(4)2 (4)2 0 t − z

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233 crit Nf m2 (Nf ) F2 (f ; Nf ) F2 (crit f ; Nf ) + + − 2 2(4)2 2f 2f (crit f )

159

;

where inside the bracket we added crit Nfcrit F2 (crit f ; Nf ) + : 0=− 2 2(4)2 (crit f )

(6.126)

(6.127)

To obtain the critical behavior we note ! l

sV t lT dz − 1 e−2z → f(j) ; t − z 4 0 Nfcrit

(Nfcrit )2 Nf → j : 2(4)2 2(4)2 2(4)2 Then Eq. (6.126) behaves as m2 (Nf ) m2 (Nf ) lT crit ∼ f(j) Nf j − f(j) + 2 2f 2f −

322 + crit Nf

(6.128)

crit F2 (f ; Nf ) F2 (crit f ; Nf ) − 2 2f (crit f )

! :

(6.129)

Since the second term in the square bracket is negative, this cannot dominate over the other terms. Thus we have to require f(j)=j1

(6.130) F2 (f ; Nf )=2f

crit crit 2 F2 (crit f ; Nf )=(f )

− in the fourth term of in Eq. (6.121). The behavior of Eq. (6.129) is determined by that of qq S through the Wilsonian matching condition (5.8). Then it is reasonable to assume that this term goes to zero faster than the 0rst term does. In addition the third term cannot dominate over the other terms, of course. As a result the critical behavior of m2 (Nf )=2f is governed by the 0rst term in the right-hand side of Eq. (6.129). This implies that m2 (Nf ) takes the form: m2 (Nf )=2f ∼ jf(j) → 0 ;

(6.131)

F2 (m (Nf ); Nf )=2f

which leads to ∼ j. The second term of RHS of Eq. (4.222) approaches to zero faster than the 0rst term does. Thus we obtain the critical behavior of the order parameter as F2 (0; Nf )=2f ∼ j → 0 : Eqs. (6.131) and (6.132) shows that m approaches to zero faster than F [106]: m2 ∼ f(j) → 0 ; F2 (0; Nf )

(6.132)

(6.133)

as we naively expected in Eq. (6.118). This is a salient feature of the VM [106]. √ Since F2 (0) is usually expected to scale as F2 (0) ∼ m2 , Eq. (6.132) implies that m ∼ j, in contrast to the essential-singularity type Eq. (6.104). This may be a characteristic feature of the

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one-loop RGEs we are using. However, the essential-singularity scaling is more sensitive to the ladder artifact than the estimate of the anomalous dimension m 1 which implies that qq S ∼ m2 . Then Eq. (6.132) implies f(j) ∼ qq S 2 ∼ m4 ∼ j2 ;

(6.134)

which will be later used as an ansatz for explicit computation of the global Nf -dependence for 3 ¡ Nf ¡ Nfcrit . 61 Let us now consider the behaviors of the physical quantities listed in Section 5.3 [see also Ref. [105]]: The – mixing strength g in Eq. (5.35) and the –– coupling constant g in Eq. (5.39) go to zero as [106] g (m ) = g(m )F2 (m ) ∼ jf1=2 (j) → 0 ; g (m ; 0; 0) =

g(m ) F2 (m ) ∼ f1=2 (j) → 0 ; 2 F2 (0)

(6.135) (6.136)

where a()=a(m )=1 was used. As discussed in Ref. [105], the KSRF (I) relation for the low-energy 2 quantities g (0) = 2g (0; 0; 0)F2 (0) holds as a low energy theorem of the HLS [23,22,103,95,96] for any Nf . The relation for on-shell quantities is violated by about 15% for Nf = 3 (see Eq. (3.73) as well as Ref. [105]). As Nf goes to Nfcrit , g (m ) and g (m ; 0; 0) approach to g (0) and g (0; 0; 0), respectively, and hence the on-shell KSRF (I) relation becomes more accurate for 2 larger Nf . On the other hand, the (on-shell) KSRF (II) relation m2 = 2g (m ; 0; 0)F2 (0) becomes 2 (m ; 0; 0)F2 (0) → 0. By substituting the less accurate. Near the critical >avor it reads as m2 = 4g critical behaviors in Eqs. (6.131), (6.135) and (6.136) into the expressions for the → decay width and the → e+ e− decay width given in Eqs. (3.64) and (3.65) with putting me = m = 0, the critical behaviors of the ratio of the width to the mass and the peak value of e+ e− → cross section are expressed as [106] 2 5=m ∼ g ∼ f(j) → 0 ;

(6.137)

2 m4 ) ∼ 1=f2 (j) → ∞ : 5ee 5 =52 ∼ g2 =(g

(6.138)

The parameters Lr10 (m ) and Lr9 (m ) de0ned in Eqs. (5.37) and (5.41) [105] diverge as Nf approaches to Nfcrit . However, we should note that, even for Nf = 3, both Lr10 ($) and Lr9 ($) have the infrared logarithmic divergences when we take $ → 0 in the running obtained by the chiral perturbation theory [79,80]. Thus we need more careful treatment of these quantities for large Nf . This is beyond the scope of this report. 61

We could also assume a case f(j) ∼ j which is a simple mean 0eld type corresponding to the NJL type scaling with m = 2. Such a behavior may be related to the following large Nf argument [104]: g is the coupling of the three-point interaction of the vector mesons. Then, as we have shown below Eq. (6.120), large Nc argument of QCD tells us that g2 behaves as 1=Nc in the large Nc limit with 0xed Nf . On the other hand, to make a large Nf expansion in the HLS consistent the gauge coupling g2 falls as 1=Nf . However, the HLS is actually related to QCD, so that the large Nf limit should be taken with Nc =Nf 0nite. This situation can be seen by rewriting the RHS of Eq. (6.121) into (gS 2 =Nc )(Nc =Nf − Nc =Nfcrit ).

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6.3.3. Nf -dependence of the parameters for 3 6 Nf ¡ Nfcrit In this subsection we illustrate how the HLS parameters would change as we vary the Nf from 3 to Nfcrit . For that purpose we need more speci0c assumption on the Nf -dependence of the QCD parameters in OPE. Here we adopt a simple ansatz which is consistent with the scaling property near the critical point given in the previous subsubsection. Let us start from the parameters of the QCD appearing in the OPE. The HLS is matched with the underlying QCD at the matching scale f . This matching scale can be regarded as the scale where the QCD running coupling becomes of order one. Thus it seems natural to require #s (f ; Nf ) to be a constant against the change of Nf . Furthermore, the large-Nc analysis shows that Nc #s (f ; Nf ) is independent of Nc . Here we show how to determine the Nf -dependence of the matching scale from this requirement. We note that theories of QCD with di/erent Nf are compared by 0xing QCD , and that it is enough to use the one-loop QCD running coupling above the matching scale since the running coupling is small at the scale above the matching scale. The one-loop running coupling is given by #($; Nf ) =

4 ; L0 (Nf ) ln($2 =2QCD )

(6.139)

where 1 L0 (Nf ) = (11Nc − 2Nf ) : 3

(6.140)

The requirement (Nc =3) #s (f ; Nf ) = constant = #s (3 ; 3)|Nc =3 0:7, with 3 = 1:1 GeV, is rewritten into the following form: Nf 3 ln(f =QCD ) = constant : L0 (Nf ) ln(f =QCD ) = 11 − 2 (6.141) Nc Nc This determines the Nf -dependence as well as the Nc -dependence of the matching scale f . Note that the Nc -dependence of the ratio f =QCD is actually very small: The di/erence between the ratio for Nc =Nf =3 and that for Nc =∞ and Nf =3 is about 2%. One might think that the Nf -dependence of the ratio f =QCD is very strong and f =QCD vanishes in the large Nf limit. However, the large Nf limit should be taken with Nf =Nc 0xed, so that the ratio f =QCD remains as constant in the large Nf limit. Actually, the ratio varies at most by 4% for 0 ¡ Nf =Nc ¡ Nfcrit =Nc 5=3. 62 62

We could use the two-loop running coupling (and the associated QCD [14]) determined by Eqs. (6.98) and (6.99) which has an infrared 0xed point for Nf ¿ Nf∗ (∼ 8 for Nc = 3) and would have more relevance to the ladder/perturbative argument which indicates Nfcrit ∼ 12Nc =3. However, our rough result Nfcrit ∼ 5Nc =3 is rather di/erent from that and is closer to the lattice result, and hence the two-loop running may not be relevant. Actually, the small dependence of f =QCD on Nc as well as Nf in the region 0 ¡ Nf =Nc ¡ 5=3 is valid even when we use the solution of the two-loop beta function in Eq. (6.141). This can be seen from the following explicit form of the solution of the two-loop beta function [12]: #∗ − #(f ; Nf ) QCD 1 1 ln = ln − ; f #(f ; Nf ) b #∗ b #(f ; Nf ) where b and #∗ are de0ned in Eqs. (6.99) and (6.100).

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As we mentioned earlier, the gluonic condensate (#s =)G$ G $ is independent of the renormalization point of QCD, so that it is reasonable to assume that it is independent of Nf , and scales as Nc [28]. So we assume # 1 " #s G$ G $ = constant : (6.142) Nc Let us now discuss the more involved estimate of the critical value Nfcrit . When we estimated the crit value of crit A in Eq. (6.113) [and then Nf in Eq. (6.114) or Eq. (6.117)], we used the same values of #s =#s (f ; Nf ), (#s =)G$ G $ and f =(Nf ) for Nf =Nfcrit as those for Nf =3. Here, although we assume that #s and (#s =)G$ G $ do not depend on Nf as in Eqs. (6.141) and (6.142), f does depend on Nf , which is determined from Eq. (6.141). Then, the critical number of >avors Nfcrit is determined by solving Nfcrit =

3(Nc2 − 1) #s 22 (#s =)G$ G $ : + 8Nc Nc 4 (Nfcrit )

(6.143)

By using #s = #s (3 = 1:1 GeV; Nf = 3) 0:69 and (#s =)G$ G $ = 0:012 GeV4 , the value of Nfcrit for Nc = 3 is estimated as 63 Nfcrit 5:0 ± 0:1 ± 0:1 ;

(6.144)

and the values of f and crit A are determined as (Nfcrit ) 1:3 ± 0:1 ± 0:01 GeV;

crit A 0:25 ± 0:03 ± 0:03 ;

(6.145)

which are compared with the previous rough estimate in Section 6.3.1: Nfcrit 5:1 (Nc =3), f =3 1:1 GeV and crit A 0:27. Now we discuss the quark condensate. As we have shown in Eq. (6.132), F2 (0) in the present approach scales as F2 (0) ∼ m2 ∼ A ≡ 1=Nf − 1=Nfcrit for any choice of the scaling property of qq S 2. On the other hand, we have argued below Eq. (6.122) that the dynamics of large Nf QCD will provide m 1 which implies qq S ∼ m2 . Then we here adopt the following ansatz for the global Nf -dependence of qq : S 1=Nf − 1=Nfcrit qq S f = : qq S 3 1=3 − 1=Nfcrit

(6.146)

Combination of Eqs. (6.141), (6.146) and (6.142) determines the Nf -dependences of the axialvector and vector current correlators derived in the OPE. Through the Wilsonian matching the Nf -dependences of the parameters in the OPE are transferred to those of the parameters in the HLS. However, as we discussed in Section 5.2, three Wilsonian matching conditions in Eqs. (5.7)–(5.9) are not enough to determine 0ve parameters F (f ; Nf ), a(f ; Nf ), g(f ; Nf ), z3 (f ; Nf ) and z2 (f ; Nf ) − z1 (f ; Nf ). As for the Nc -dependence of the HLS gauge coupling g, as we discussed below Eq. (6.121), g2 scales as 1=Nc . Then, from Eq. (6.119) together with the assumption that each 63

The center values in Eqs. (6.144) and (6.145) are given for (3 ; QCD ) = (1:1; 0:4) GeV, and the 0rst and second errors are obtained by allowing 3 and QCD to vary 3 = 0:1 GeV and QCD = 0:05 GeV, respectively.

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163

term in the left-hand side have the same scaling property, we see that z3 scale as Nc . Then we use the following assumptions for the Nc - and Nf -dependences of g(; Nf ) and z3 (; Nf ): 2 1=Nf − 1=Nfcrit g2 (f ; Nf ) = ; (6.147) g2 (3 ; 3) 1=3 − 1=Nfcrit 1 z3 (f ; Nf ) = constant : (6.148) Nc Note that the condition in Eq. (6.147) is consistent with Eq. (6.134) or Eq. (6.146) through the condition in Eq. (6.120). From the above assumptions we can determine the Nf -dependences of other three bare parameters through the Wilsonian matching. Now that we have determined the Nf -dependences of 0ve parameters F (f ; Nf ), a(f ; Nf ), g(f ; Nf ), z3 (f ; Nf ) and z2 (f ; Nf ) − z1 (f ; Nf ) in the HLS. we study the Nf -dependences of the physical quantities by solving the RGEs with Nc = 3 0xed. To determine the current correlators in the OPE for Nf = 3 we use # "# s G$ G $ = 0:012 GeV4 ; qq S 1 GeV = −(0:225 GeV)3 ;

(6.149)

as a typical example. To determine the parameters in the HLS for Nf = 3 through the Wilsonian matching we use 3 = 1:1 GeV ;

QCD = 400 MeV ;

(6.150)

for illustration. First, in Fig. 18, we show the Nf -dependences of F (f ; Nf )=f and a(f ; Nf ) together with those of [a(f ; Nf ) − 1]=g2 (f ; Nf ) and [z2 (f ; Nf ) − z1 (f ; Nf )]=g2 (f ; Nf ) which are determined through the Wilsonian matching conditions (5.8), (5.9) and (5.7) together with the above assumptions of the Nf -dependences of other parameters. Fig. 18(a) shows that the ratio F (f ; Nf )=f has only small Nf -dependence as we have discussed before. From Fig. 18(b) we can see that the value of a(f ; Nf ) is close to one in most region. Figs. 18(c) and (d) show that a(f ; Nf ) − 1 and z2 (f ; Nf ) − z1 (f ; Nf ) actually scale as g2 (f ; Nf ) and qq S 2 near the critical >avor Nfcrit 5 as we have discussed below Eq. (6.119). Next, we show the Nf -dependences of F (0; Nf )=f and m (Nf )=f in Fig. 19. This shows that F (0; Nf ) and m (Nf ) smoothly go to zero when Nf → Nfcrit . 64 Next we show in Fig. 20 the Nf -dependences of g , g and a(0; Nf ) which were de0ned in Section 5.3. The Nf -dependence of a(0) shows that the vector dominance is already largely violated even o/ the critical point. Finally, to check the KSRF relations I and II in large Nf QCD [see Section 3.5], we show the Nf -dependences of g =(2g F2 (0)) [ = 1 − g2 (m )z3 (m )] and m2 =(2g F2 (0)) [ = 2=a(0)] in Fig. 21, the unity value of which corresponds to the KSRF relations. This shows that the KSRF I relation, which is the low energy theorem of the HLS, approaches to the exact relation near the critical point, while the KSRF II relation is largely violated there as they should (due to the VM; a(0) → 1, g2 (m ) → 0). 64

In Fig. 19, the value of m (Nf )=f becomes small already at the o/-critical point. This is due to the ansatz of Nf -dependence of g2 (f ; Nf ) adopted in Eq. (6.147). If we used the ansatz of essential-singularity-type scaling suggested by the Schwinger–Dyson approach [14,12,148], on the other hand, the mass m (and other physical quantities as well) would not change much o/ the critical point but suddenly approach the critical point value only near the critical point.

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0.14

1.4

0.12

1.2

0.1

a (Λf ;N f)

0.08 0.06 0.04

0.8 0.6 0.4

0.02

0.2 3.5

4

(a)

5

4.5

3.5

Nf 0.025 0.02 0.015 0.01 0.005 3.5

4

(c)

4

(b)

4.5

5

5

4.5

Nf [z2 (Λf ;N f) − z1 (Λf ;N f)]

[a (Λf ;N f) − 1] / g2 (Λf ;N f)

1

-0.015

/ g2 (Λf ;N f)

Fπ(Λf ;N f) / Λf

164

-0.0175 -0.02 -0.0225 -0.025 -0.0275 3.5

4

5

4.5

Nf

(d)

Nf

Fig. 18. Nf -dependences of (a) F (f ; Nf )=f , (b) a(f ; Nf ), (c) [a(f ; Nf ) − 1]=g2 (f ; Nf ) and (d) [z2 (f ; Nf ) − z1 (f ; Nf )]=g2 (f ; Nf ).

0.7 0.6

0.06

mρ(N f) / Λf

Fπ(0;N f) / Λf

0.08

0.04

0.5 0.4 0.3 0.2

0.02

0.1

3.5 (a)

4

4.5

Nf

5

3.5

4

(b)

Fig. 19. Nf -dependences of (a) F (0; Nf )=f and (b) m (Nf )=f .

4.5

Nf

5

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7

0.1

6 5

gρππ(N f)

f

gρ(N f) / Λ

2

0.08 0.06 0.04

4 3 2

0.02

1

3.5

4

(a)

5

4.5

3.5

4

(b)

Nf

5

4.5

Nf

2

a(0;N f)

1.5 1 0.5

3.5

4

(c)

5

4.5

Nf

Fig. 20. Nf -dependences of (a) g , (b) g and (c) a(0; Nf ).

1.4

2

0.6 0.4

2 2gρππ Fπ(0)

1 0.8

mρ2

gρ

2

2gρππ Fπ(0)

1.2 1.5 1 0.5

0.2 3.5

(a)

4

Nf

4.5

5

3.5

(b)

4

4.5

5

Nf

Fig. 21. Nf -dependences of KSRF relations (I) and (II): (a) g =(2g F2 (0)) and (b) m2 =(2g F2 (0)).

6.3.4. Vector dominance in large Nf QCD Since Sakurai advocated vector dominance (VD) as well as vector meson universality [165], VD has been a widely accepted notion in describing vector meson phenomena in hadron physics. In fact several models such as the gauged sigma model (see, e.g., Refs. [127,141]) are based on VD to

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introduce the photon 0eld into the Lagrangian. Moreover, it is often taken for granted in analysing the dilepton spectra to probe the phase of quark–gluon plasma for the hot and/or dense QCD (see, e.g., Refs. [159,130,162]). As far as the well-established hadron physics for the Nf = 3 case is concerned, it in fact has been extremely successful in many processes such as the electromagnetic form factor of the pion [165] and the electromagnetic transition form factor (see, e.g., Ref. [31]), etc. as studied in Section 3.8. However, there has been no theoretical justi0cation for VD and as it stands might be no more than a mnemonic useful only for the three->avored QCD at zero temperature/density. Actually, as studied in Section 3.8, VD is already violated for the three->avored QCD for the anomalous processes such as → 3=0 → 2 [74,24] and ! transition form factor (see, e.g., Ref. [40,18,19]). This strongly suggests that VD may not be a sacred discipline of hadron physics but may largely be violated in the di/erent parameter space than the ordinary three->avored QCD (non-anomalous processes) such as in the large Nf QCD, Nf being number of massless >avors, and hot and/or dense QCD where the chiral symmetry restoration is expected to occur. It is rather crucial whether or not VD is still valid when probing such a chiral symmetry restoration through vector meson properties [158,42–45]. Here we emphasize that in the hidden local symmetry (HLS) model [21,24] the vector mesons are formulated precisely as gauge bosons; nevertheless VD as well as the universality is merely a dynamical consequence characterized by the parameter choice a = 2 (see Section 3.5). In this section we study the vector dominance (VD) in large Nf QCD following Ref. [107]. Here it is convenient to use the parameters X ($) and G($) de0ned in Eqs. (4.254) and (4.255). The VD is characterized by a(0)=2, where a(0) is de0ned in Eq. (5.44). Substituting Eqs. (4.219) and (4.220) with Eq. (4.254) into Eq. (5.44), we obtain a(0) = a(m )=[1 + a(m )X (m ) − 2X (m )] :

(6.151)

This implies that the VD (a(0)=2) is only realized for (X (m ); a(m ))=(1=2; any) or (any, 2) [107]. In Nf = 3 QCD, the parameters at m scale, (X (m ); a(m ); G(m )) (0:51; 1:38; 0:37), happen to be near such a VD point. However, the RG >ow actually belongs to the 0xed point (X4∗ ; a∗4 ; G4∗ ) which is far away from the VD value. Thus, the VD in Nf = 3 QCD is accidentally realized by X (m ) ∼ 1=2 which is very unstable against the RG Dow [107] (see Fig. 14). For G = 0 (Fig. 12) the VD holds only if the parameters are (accidentally) chosen to be on the RG >ow entering (X; a; G) = (0; 2; 0) which is an end point of the line (X (m ); a(m )) = (any; 2). For a = 1 (Fig. 13), on the other hand, the VD point (X; a; G) = (1=2; 1; 1=2) lies on the line (X (m ); a(m )) = (1=2; any). Then, phase diagrams in Figs. 12 and 13 and their extensions to the entire parameter space (including Fig. 14) show that neither X (m ) = 1=2 nor a(m ) = 2 is a special point in the parameter space of the HLS. Thus the VD as well as the universality can be satis?ed only accidentally [107]. Therefore, when we change the parameter of QCD, the VD is generally violated. In particular, neither X (m ) = 1=2 nor a(m ) = 2 is satis0ed on the phase boundary surface characterized by Eq. (4.265) where the chiral restoration takes place in HLS model. Therefore, VD is realized nowhere on the chiral restoration surface [107]. Moreover, when the HLS is matched with QCD, only the point (X2∗ ; a∗2 ; G2∗ ) = (1; 1; 0), the VM point, on the phase boundary is selected, since the axialvector and vector current correlators in HLS can be matched with those in QCD only at that point [106]. Therefore, QCD predicts a(0) = 1, i.e., large violation of the VD at chiral restoration. Actually, as is seen in Fig. 20(c), for the chiral

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Table 17 Duality and conformal window in N = 1 SUSY QCD Nf

“Electric theory” SU(Nc ) SQCD

“Magnetic theory” SU(Nf − Nc ) SQCD

↑ 3Nc

Free non-Abelian electric theory IR free

Strong no-Abelian magnetic theory Asymptotic free

3Nc =2

(Interacting non-Abelian Coulomb phase) IR 0xed point

IR 0xed point

Nc + 2

Strong non-Abelian electric theory Asymptotic free

Free non-Abelian magnetic theory IR free

Nc + 1

complete con0nement No S,SB (s-con0nement)

completely Higgsed

Nc

complete con0nement S,SB

completely Higgsed

restoration in the large Nf QCD [118,14] the VM can in fact takes place [106], and thus the VD is badly violated [107]. 6.4. Seiberg-type duality Increasing attention has been paid to the duality in various contexts of modern particle theory. Seiberg found the “electric-magnetic” duality in N = 1 Supersymmetric (SUSY) QCD with Nc colors and Nf >avors [170]. The Nf -dependence of the theory is summarized in Table 17. For the region 32 Nc ¡ Nf ¡ 3Nc (“conformal window”) in the SUSY QCD, there exists a “magnetic theory” with the SU(Nf − Nc ) gauge symmetry which is dual to the original SU(Nc ) theory regarded as the “electric theory”. Although the origin of the magnetic gauge symmetry (“induced at the composite level”) is not obvious from the original theory, both theories in fact have the infrared (IR) 0xed point with exact conformal symmetry and with the same IR physics. This region is called “interacting non-Abelian Coulomb phase”. When Nf decreases, the electric theory becomes stronger in IR, while the magnetic theory gets weaker, with the magnetic gauge group being reduced through the Higgs mechanism. Decreasing Nf further beyond the conformal window, we 0nally arrive at Nf =Nc where the magnetic theory is in complete Higgs phase (reduced to no gauge group), which corresponds to the complete con0nement (and spontaneously broken chiral symmetry) of the electric theory. Similar conformal window may also exist in the ordinary (non-SUSY) QCD with massless Nf >avors (33Nc =2 ¿ Nf ¿ Nfcrit ∼ 5(Nc =3)), as was discussed in Section 6.2. Situation including the proposal in Ref. [104] is summarized in Table 18. Here we recall that, for small Nf , the vector mesons such as the meson can be regarded as the dynamical gauge bosons of HLS [21,24]. The HLS is completely broken through the Higgs mechanism as the origin of the vector meson mass. This gauge symmetry is induced at the composite level and has nothing to do with the fundamental color gauge symmetry. Instead, the HLS is associated with the >avor symmetry.

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Table 18 Duality and conformal window (33Nc =2 ¿ Nf ¿ Nfcrit ∼ 5(Nc =3)) in QCD Nf

“Electric Theory” SU(Nc ) QCD

“Magnetic Theory” SU(Nf ) HLS

↑ 11Nc =2

Free electric theory IR free

EFT ?

11Nc =2

Interacting non-Abelian Coulomb phase IR 0xed point (No S,SB/Con0nement)

EFT ?

∼ 5(Nc =3) Nc

Conformal phase transition Con0ned electric theory (S,SB) “real world”(SU(3) QCD)

Vector Manifestation Higgsed magnetic theory (S,SB) SU(3) HLS

In Ref. [104] we found that the Seiberg duality is realized also in the ordinary (non-SUSY) QCD through the HLS. For small Nc (=3) 6 Nf ¡ Nfcrit ∼ 5(Nc =3), the SU(Nf ) HLS is in complete Higgs phase and yields the same IR physics as the SU(Nc ) QCD in the con0nement/chiral-symmetrybreaking phase, and plays the role of the “Higgsed magnetic gauge theory” dual to the “Con0ned electric gauge theory” (QCD) in the spirit of Seiberg duality. Then the mesons can in fact be regarded as the Higgsed “magnetic gluons” of the SU(Nf ) HLS. In order for such a duality between QCD and the HLS be consistently satis0ed, there should be a way that the chiral restoration takes place for large Nf also in the HLS theory by its own dynamics. We have already seen in Section 6.3 that the HLS can provide the chiral restoration by its own dynamics for a certain value of Nf = Nfcrit 5(Nc =3) which is in rough agreement with 6 ¡ Nfcrit ¡ 7 found in the lattice simulation of the electric theory, the QCD with Nc = 3. Thus the Seiberg-type duality does exist also in the ordinary (non-SUSY) QCD at least for Nc (=3) 6 Nf ¡ Nfcrit ∼ 5(Nc =3) [104]. We do not know at this moment, however, what the duality would be for 11Nc =2 ¿ Nf ¿ Nfcrit where the EFT like HLS may not exist because of a possible absence of e/ective 0elds of light bound states in the symmetric phase, as was suggested by the conformal phase transition (see Section 6.1.3). It should also be emphasized that near the critical point this Higgsed magnetic gauge theory prodides an example of a weakly-coupled composite gauge theory with light gauge boson and NG boson, while the underlying electric gauge theory is still in the strongly coupled phase with con0nement and chiral symmetry breaking. This unusual feature may be useful for model building beyond the Standard Model.

7. Renormalization at any loop order and the low energy theorem As was discussed in Section 3, the KSRF relation (version I) [see Eq. (3.62)], g = 2F2 g ;

(7.1)

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holds as a “low energy theorem” of the HLS [23], which was 0rst proved at the tree level [22], then at one-loop level [103] and further at any loop order [95,96]. In this section we brie>y review the proof of the low energy theorem of the HLS at any loop order, following Refs. [95,96]. Although Refs. [95,96] presumed only logarithmic divergence, only the relevant assumption made there was that there exists a symmetry preserving regularization. As was discussed in Section 4.5, inclusion of the quadratic divergence through the replacement in Eq. (4.85) is in fact consistent with the gauge invariance. Then the proposition and the proof below are valid even if we include the quadratic divergences. We restrict ourselves to the chiral symmetric case, 65 so that we take ,ˆ = 0 in the leading order Lagrangian in Eq. (4.20): L(2) = LA + aLV + Lkin (V$ ) ;

(7.2)

where LA and aLV are de0ned in Eqs. (3.33) and (3.34), respectively: LA ≡ F2 tr[#ˆ⊥$ #ˆ$⊥ ] ;

(7.3)

aLV ≡ F2 tr[#ˆ $ #ˆ$ ] :

(7.4)

It should be noticed that in this section we classify LA and LV as “dimension-2 terms” and Lkin as “dimension-4 term”, based on counting the dimension of only the ?elds and derivatives. This is somewhat di/erent from the chiral counting explained in Section 4.1 where the HLS gauge coupling carries O(p), and thus Lkin is counted as O(p2 ). The counting method adopted in this section is convenient for classifying the terms with the same chiral order: The contribution at nth loop order is expected to generate O(p2n+2 ) corrections which take the form of (g2 )n LA , (g2 )n LV , (g2 )n Lkin , and so on. In (g2 )n LA and (g2 )n LA , the O(p2 ) out of O(p2n+2 ) is carried by derivatives and 0elds, while in (g2 )n Lkin , the O(p4 ) out of O(p2n+2 ) is by them. Then, by counting the dimensions of only the 0elds and derivatives, we can extract the terms relevant to the low-energy region out of all the possible n-loop corrections. Note that we focus on the renormalizability of the terms of dimension two, LA and LV terms in Eq. (7.2), which is just what we need for proving the low energy theorem. We introduce the BRS transformation and make the proposition in Section 7.1. We prove the proposition in Section 7.2. Finally, in Section 7.3, we prove that the low-energy theorem in Eq. (7.1) holds at any-loop order. Also note that, in this section, we use the covariant gauge instead of the background 0eld gauge, since the higher order loop calculation is well-de0ned compared with the background 0eld gauge. Also the o/-shell extrapolation is easily done in covariant gauge compared with the RC gauge (see Section 7.3). 7.1. BRS transformation and proposition Let us take a covariant gauge condition for the HLS, and introduce the corresponding gauge-0xing and Faddeev–Popov (FP) terms: LGF + LFP = Ba 9$ V$a + 12 #Ba Ba + iCS a 9$ D$ C a ; (7.5) 65

See Refs. [95,96] for the e/ect of the symmetry breaking mass terms of NG 0elds.

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

where Ba is the Nakanishi–Lautrap (NL) 0eld and C a (CS a ) the FP ghost (anti-ghost) 0eld. As in the previous sections we do not consider the radiative corrections due to the external gauge 0elds V$i ≡ (La$ ; R$a ), so that we need not introduce the gauge-0xing terms for V$ . Then, the corresponding ghost 0elds Ci ≡ (CLa ; CRa ) are non-propagating. The in0nitesimal form of the Gglobal × Hlocal transformation (3.2) is given by C(x) = i:(x)C(x) − iC(x)#(x) ; :(x) ≡ : a (x)Ta ;

#(x) ≡ #a (x)Ta :

(7.6)

This de0nes the transformation of the Nambu–Goldstone (NG) 0eld =i ≡ (a =F ; a =F ) [see Eq. (3.4)] in the form =i = : a Wai (=) + #j Wji (=)(≡ : A WAi (=)) ;

(7.7)

where A denotes a set (a; i) of labels of Hlocal and Gglobal . Accordingly, the BRS transformation of the NG 0elds =i , the gauge 0elds V$A ≡ (V$a ; V$i ) and the FP ghost 0elds C A ≡ (C a ; Ci ) are respectively given by 9 i A ˆ i i ˆ B = = C WA = WA ≡ WA (=) i ; 9= B V$A = 9$ C A + V$B C C fBC A ; 1 B C A = − C B C C fBC A : 2

(7.8)

For de0niteness we de0ne the dimension of the 0elds as dim[=i ] = 0;

dim[V$A ] = 1 :

(7.9)

It is also convenient to assign the following dimensions to the FP-ghosts: dim[C A ] = 0;

dim[CS a ] = 2 :

(7.10)

Then the BRS transformation does not change the dimension. According to the above dimension counting, we may divide the Lagrangian Eq. (7.2) plus Eq. (7.5) into the following two parts: (a) dimension-2 part LA + aLV , (b) dimension-4 part Lkin (V$ ) + LGF + LFP , where we count the dimension of the 0elds and derivatives only. Now, we consider the quantum correction to this system at any loop order, and prove the following proposition. Proposition. As far as the dimension-2 operators are concerned, all the quantum corrections, including the ?nite parts as well as the divergent parts, can be absorbed into the original dimension-2 Lagrangian LA + aLV by a suitable rede?nition (renormalization) of the parameters a, F2 , and the ?elds =i , V$a .

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This implies that the tree-level dimension-2 Lagrangian, with the parameters and 0elds substituted by the “renormalized” ones, already describes the exact action at any loop order, and therefore that all the “low energy theorems” derived from it receive no quantum corrections at all. 7.2. Proof of the proposition We prove our proposition in the same way as the renormalizability proof for gauge theories [30] and two-dimensional nonlinear sigma models [37,38]. We can write down the WT identity for the e/ective action 5. The NL 0elds Ba and the FP anti-ghost 0elds CS a can be eliminated from 5 by using their equations of motion as usual. Then the tree level action S = 5tree reads S[P; K; a] = S2 [=; V ] + S4 [P; K] ;

S2 [=; V ] = d 4 x (a⊥ LA (=; V ) + a LV (=; V )) ;

d 4 x (Lkin (V$ ) + K · B P) ;

S4 [P; K] =

(7.11)

where P ≡ (=i ; V$A ; C A ) are the 0eld variables and K ≡ (Ki ; KA$ ; LA ) (KA$ ≡ (Ka$ ; K$i ), LA ≡ (La ; Li )) denote the BRS source 0elds for the NG 0eld =i , the gauge 0elds V$A and the ghost 0elds C A , respectively; i.e., KA$ B V$A = Ka$ B V$a + K$i V$i ; L A B C A = L a B C a + L i B C i : We have rewritten

F2

a⊥ f2 ≡ F2 ;

and

F2

(7.12)

as

a f2 ≡ F2 ;

(7.13)

so that the renormalization of F2 and F2 corresponds to that of a ≡ (a ; a⊥ ). According to the dimension assignment of the 0elds, the dimension of the above BRS source 0elds K is given by dim[Ki ] = dim[LA ] = 4;

dim[KA$ ] = 3 :

(7.14)

The WT identity for the e/ective action 5 is given by 5∗5=0 ;

(7.15)

where the ∗ operation is de0ned by ←

←

F G F G − (−)P F ∗ G = (−) P K K P P

(7.16) ←

for arbitrary functionals F[P; K] and G[P; K]. (Here the symbols and denote the derivatives from the left and right, respectively, and (−)P denotes +1 or −1 when P is bosonic or fermionic, respectively.) The e/ective action is calculated in the loop expansion: 5 = S + ˝5(1) + ˝2 5(2) + · · · :

(7.17)

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The ˝n term 5(n) contains contributions not only from the genuine n-loop diagrams but also from the lower loop diagrams including the counter terms. We can expand the nth term 5(n) according to the dimension: 5(n) = 50(n) [=] + 52(n) [=; V ] + 54(n) [P; K] + · · · :

(7.18)

Here again we are counting the dimension only of the 0elds and derivatives. The 0rst dimension-0 term 50(n) can contain only the dimensionless 0eld =i without derivatives. The two dimensions of the second term 52(n) is supplied by derivative and/or the gauge 0eld V$A . The BRS source 0eld K carries dimension 4 or 3, and hence it can appear only in 54(n) and beyond: the dimension-4 term 54(n) is at most linear in K, while the dimension-6 term 56(n) can contain a quadratic term in Ka$ , the BRS source of the hidden gauge boson V$a . To calculate 5(n) , we need to use the “bare” action, (S0 )n = S[(P0 )n ; (K0 )n ; (a0 )n ] ;

(7.19)

where the nth loop order “bare” 0elds (P0 )n , (K0 )n and parameters (a0 )n are given by (P0 )n = P + ˝P(1) + · · · + ˝n P(n) ; (K0 )n = K + ˝K(1) + · · · + ˝n K(n) ; (a0 )n = a + ˝a(1) + · · · + ˝n a(n) :

(7.20)

Let us now prove the following by mathematical induction with respect to the loop expansion parameter n: (I) 50(n) (=) = 0. (II) By choosing suitably the nth order counter terms P(n) , K(n) and a(n) , 52(n) [=; A] and the K-linear terms in 54(n) [P; K] can be made vanish; 52(n) [=; V ] = 54(n) [P; K] =0 : K-linear (III) The 0eld reparameterization (renormalization) (P; K) → ((P0 )n ; (K0 )n ) is a “canonical” transformation which leaves the ∗ operation invariant. Suppose that the above statements are satis0ed for the (n − 1)th loop order e/ective action 5(n−1) . We calculate, for the moment, the nth loop e/ective action 5(n) using the (n − 1)th loop level “bare” action (S0 )n−1 , i.e., without nth loop counter terms. We expand the ˝n terms in the WT identity n− 1 1 (l) S ∗ 5(n) = − 5 ∗ 5(n−l) ; (7.21) 2 l=1

according to the dimensions like in Eq. (7.18). Then using the above induction assumption, we 0nd: S4 ∗ 50(n) + S2 ∗ 52(n) = 0 (dim 0) ;

(7.22)

S4 ∗ 52(n) + S2 ∗ 54(n) = 0 (dim 2) ;

(7.23)

S4 ∗ 54(n) + S2 ∗ 56(n) = 0 (dim 4) :

(7.24)

These three renormalization equations give enough information for determining possible forms of 50(n) , 52(n) and 54(n) |K-linear (the K-linear term in 54(n) ) which we are interested in.

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Noting that the BRS transformation B on the 0elds P ≡ (=i ; V$A ; C A ) can be written in the form B =

S4 ; K P

(7.25)

we see it convenient to de0ne an analogous transformation 5 on the 0elds P by 5 ≡

54(n) : K P

(7.26)

Then we can write 54(n) in the form 54(n) = A4 [=; V ] + Ki 5 =i + KA$ 5 V$A + LA 5 C A :

(7.27)

In terms of this notation, Eqs. (7.22)–(7.24) can be rewritten into B 50(n) = 0 ;

(7.28)

B 52(n) + 5 S2 = 0 ;

(7.29)

56(n) S2 =0 : (7.30) K P First, let us consider the dimension-0 part of the renormalization equation (7.28). Since there are no invariants containing no derivatives, we can immediately conclude 50(n) = 0, and hence our statement (I) follows. Next, we consider the dimension-2 and dimension-4 parts of the renormalization equations (7.29) and (7.30). A tedious but straightforward analysis [95,96] of the K-linear term in Eq. (7.30) determines the general form of the 54(n) |K-linear and 56(n) |K-quadratic terms: the solution for 54(n) |K-linear or equivalently 5 is given by B 54(n) + 5 S4 +

5 C a = LB C a ;

(7.31)

i ˆ j ; F]}= ˆ + LWˆ a ) + Cj [W ˆ 5 =i = {C a ([Wˆ a ; F] ;

(7.32)

˜ a$ ) ; 5 V$a = #9$ C a + LB V$a + B (V$a − V

(7.33)

where #, L and are constants, ˜ $ ≡ CL L$ C†L − i9$ CL · C†L + CR R$ C†R − i9$ CR · C†R ; V

(7.34)

and Fˆ ≡ F i (=)9=9=i , with F i (=) being a certain dimension-0 function. Note that 5 V$i = 5 Ci = 0, since the external Gglobal -gauge 0elds V$i and their ghosts Ci are not quantized and hence their BRS source 0elds K$i and Li appear only in the tree action. Using 5 thus obtained, we next solve the above WT identity (7.29) and easily 0nd (n) a ˆ 2 + #V$ 52 = A2GI [=; V ] − FS (7.35) S2 ; V$a where A2GI is a dimension-2 gauge-invariant function of =i and V$A .

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The solutions are combined into a simple form 52(n) + 54(n) = A2GI [=; V ] − S ∗ Y (7.36) K-linear up to irrelevant terms (dimension-6 or K-independent dimension-4 terms), where the functional Y is given by

(7.37) Y = d 4 x[Ki F i (=) + #Ka$ V$a + LLa C a + fabc Ka$ Kb$ C c ] : Now, we prove our statements (II) and (III) in the above. We have calculated the above e/ective action 5(n) without using nth loop level counter terms P(n) , K(n) and a(n) . If we include those, we have the additional contribution given by S S 9S + K(n) + a(n) ; (7.38) _5(n) = P(n) P K 9a where S[P; K; a] is the tree-level action. So the true nth loop level e/ective action is given by (n) 5(n) + _5(n) ≡ 5total :

(7.39)

The tree-level action S2 is the most general gauge-invariant dimension-2 term, so that A2GI [=; V ] term in Eq. (7.36) can be canceled by suitably chosen counter terms, a(n) 9S=9a. The second term −S ∗ Y term in Eq. (7.36) just represents a “canonical transformation” of S generated by −Y . Therefore we choose the nth order 0eld counter terms P(n) and K(n) to be equal to the canonical transformations of P and K generated by +Y ; P(n) = P ∗ Y;

K(n) = K ∗ Y :

(7.40)

Then the 0rst and the second terms in Eq. (7.38) just give S ∗ Y and precisely cancel the second term in Eq. (7.36). Thus we have completed the proof of our statements (II) and (III). 7.3. Low energy theorem of the HLS In the previous subsections% of this section, we have shown in the covariant gauges that our tree-level dimension-2 action d 4 x(LA + aLV ), if written in terms of renormalized parameters and 0elds, already gives the exact action 52 including all the loop e/ects. This form of the e/ective action (in particular the LV part) implies that the previously derived relation [23,22] gV (p2 ) = 2F2 (7.41) gV (p2 ; p2 = p2 = 0) 1

2

p2 =0

is actually an exact low energy theorem valid at any loop order. Of course, this theorem concerns o/-shell quantities at p2 = 0, and hence is not physical as it stands. However, as discussed in Section 4 (see also Refs. [85,86]), we can perform the systematic low energy expansion in the HLS when the vector meson can be regarded as light. We expect that the on-shell value of gV =gV at p2 = m2V can deviate from the LHS of Eq. (7.41) only by a quantity of order m2 =2, , since the contributions of the dimension-4 or higher terms in the e/ective action 5 (again representing all the loop e/ects) are suppressed by a factor of p2 =2, at least. Therefore as far as the vector mass is light, our theorem is truly a physical one. In the actual world of QCD, the meson mass is not so light (m2 =2, ∼ 0:5) so that the situation becomes a bit obscure. Nevertheless, the fact that the

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KSRF (I) relation g =g = 2F2 holds on the mass shell with good accuracy strongly suggests that the meson is the hidden gauge 0eld and the KSRF (I ) relation is a physical manifestation of our low energy theorem. Our conclusion in this section remains unaltered even if the action S contains other dimension-4 or higher terms, as far as they respect the symmetry. This is because we needed just (S ∗ 5)2 and (S ∗ 5)4 |K-linear parts in the WT identity to which only S2 and K-linear part of S4 can contribute. When we regard this HLS model as a low energy e/ective 0eld theory of QCD, we must take account of the anomaly and the corresponding Wess–Zumino–Witten term 5WZW . The WT identity now reads 5 ∗5 =(anomaly). However, the RHS is saturated already at the tree level in this e/ective Lagrangian and so the WT identity at loop levels, which we need, remains the same as before. The WZW term 5WZW or any other intrinsic-parity-odd terms [74] in S are of dimension-4 or higher and hence do not change our conclusion as explained above. Since the low energy theorem concerns o/-shell quantities, we should comment on the gauge choice. In the covariant gauges which we adopted here, the Gglobal and Hlocal BRS symmetries are separately preserved. Accordingly, the V$ 0eld is multiplicatively renormalized (recall that V$(n) = V$ ∗ Y = #V$ ), and the above (o/-shell) low energy theorem (7.41) holds. However, if we adopt RC -gauges (other than Landau gauge), these properties are violated; for instance, =9$ = or the external gauge 0eld V$ gets mixed with our V$ through the renormalization, and our o/-shell low energy theorem (7.41) is violated. This implies that the V$ 0eld in the RC gauge generally does not give a smooth o/-shell extrapolation; indeed, in RC gauge with gauge parameter # ≡ 1=C, the correction to g =g by the extrapolation from p2 = m2 to p2 = 0 is seen to have a part proportional to #g2 =162 , which diverges when # becomes very large. Thus, in particular, the unitary gauge [see Section 3.3], which corresponds to # → ∞, gives an ill-de0ned o/-shell 0eld. Our argument is free from infrared divergences at least in Landau gauge. This can be seen as follows. In this gauge the propagators of the NG bosons, the hidden gauge bosons and the FP S are all proportional to 1=f2 in the infrared ghosts (after rescaling the FP anti-ghost CS into f2 C) region. Therefore, a general L-loop diagram, which includes V4 dimension-4 vertices and K BRS source vertices, yields an amplitude proportional to (1=f2 )(L−1+V4 +K) [190]. Thus, from dimensional consideration we see that there is no infrared contribution to 50(n) [=], 52(n) [=; V ] and 54(n) [P; K]|K-linear . In other covariant gauges, there appears a dipole ghost in the vector propagator, which is to be de0ned by a suitable regularization.

8. Towards hot and/or dense matter calculation In this section we consider an application of the approach introduced in this report to the hot and/or dense matter calculation. In hot and/or dense matter, the chiral symmetry is expected to be restored (for reviews, see, e.g., Refs. [109,160,43,111,194,162,45]). The BNL Relativistic Heavy Ion Collider (RHIC) has started to measure the e/ects in hot and/or dense matter. One of the interesting quantities in hot and/or dense matter is the change of -meson mass. In Refs. [42,43] it was proposed that the -meson mass scales like the pion decay constant in hot and/or dense matter, and vanishes at the chiral phase transition point.

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The vector manifestation (VM) reviewed in Section 6 is a general property in the chiral symmetry restoration when the HLS can be matched with the underlying QCD at the critical point. In Ref. [106] the application of the VM to the large Nf chiral restoration was done. It was then suggested [106,44,45] that the VM can be applied to the chiral restoration in hot and/or dense matter. Recently, it was shown the VM actually occurs in hot matter at zero density [99] and also in dense matter at zero temperature [93]. The purpose of this section is to give an outline of the application of the chiral perturbation, the Wilsonian matching and the VM of the HLS to the hot and/or dense matter calculation based on these works. We 0rst consider the hot matter calculation at zero density. In the low temperature region, the temperature dependence of the physical quantities are expected to be dominated by the hadronic thermal e/ects. Inclusion of the hadronic thermal corrections to the -meson mass within the framework of the HLS has been done by several groups (see, e.g., Refs. [134,174,102,162]). However, most of them included only the thermal e/ect of pions and dropped the thermal e/ects of the meson itself. In Ref. [102], the 0rst application of the systematic chiral perturbation reviewed in the previous sections to the hot matter calculation was made. There hadronic thermal e/ects were included, based on the systematic chiral perturbation in the HLS, by calculating the one-loop corrections in hot matter in the Landau gauge. We review the chiral perturbation of the HLS in hot matter following Ref. [102] in Section 8.1. A part of the calculation in the background 0eld gauge which we introduced in Section 4 is shown in Ref. [94], and the complete version will be shown in Ref. [100]. In Ref. [99], an application of the Wilsonian matching explained in Section 5 to hot matter calculation was done. In Section 8.2, we brie>y review the analysis. The main result in Ref. [99] is that by imposing the Wilsonian matching of the HLS with the underlying QCD at the critical temperature, where the chiral symmetry restoration takes place, the vector manifestation (VM ) necessarily occurs: The vector meson mass becomes zero. Accordingly, the light vector meson gives a large thermal correction to the pion decay constant, and the value of the critical temperature becomes larger than the value estimated by including only the pion thermal e/ect. The result that the vector meson becomes light near the critical temperature is consistent with the picture shown in Refs. [42–45]. In Section 8.3 we brie>y review an application of the Wilsonian matching and the VM to dense matter calculation recently done in Ref. [93]. It was shown that the VM is realized in dense matter at the chiral restoration with the mass m going to zero at the critical point. To avoid confusion, we use f (T ) [f ( $)] 66 for the physical decay constant of at non-zero temperature [density], and F for the parameters of the Lagrangian. Similarly, M denotes the parameter of the Lagrangian and m the pole mass. 8.1. Hadronic thermal eAects In this subsection we show the hadronic thermal corrections to the pion decay constant and the vector meson mass following Ref. [102], where the calculation was performed in the Landau gauge with the ordinary quantization procedure. 66

In Ref. [93] $ is used for expressing the chemical potential. Throughout this report, however, we use $ for the energy for the chemical potential. scale, and then we use $

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177

In Ref. [102] the pion decay constant at non-zero temperature was de0ned through the axialvector current correlator following the de0nition given in Ref. [39]. In Eq. (5.1), two-point function of the a axialvector current J5$ is expressed by one tensor structure. At non-zero temperature, however, we can decompose this current correlator into longitudinal and transverse pieces as

a b i d 4 x eipx 0 | TJ5$ (x)J5 (0)| T = ab [PT$ GAT (p0 ; p ˜ ; T ) + PL$ GAL (p0 ; p ˜ ; T )] ; (8.1) where polarization tensors PL and PT are de0ned in Eq. (A.42) in Appendix A.5. It is natural to de0ne the pion decay constant at non-zero temperature through the longitudinal component in the low energy limit [39]: 67 f2 (T ) ≡ − lim GAL (p0 ; p ˜ = 0; T ) : p0 →0

(8.2)

There are two types of contributions to this Green function in the HLS: (i) the pion exchange diagrams, and (ii) the contact or one-particle irreducible (1PI) diagrams. The contribution (i) is a - coupling can be corrected proportional to p$ or p at one loop. At most only one of the J5$ at one loop, which is not generally proportional to the four-momentum p$ . The other coupling is the tree-level one and proportional to p$ . When we act with the projection operator PL$ , the term proportional to p$ vanishes. Because of the current conservation we have the same kinds of contributions from 1PI diagrams: Those are roughly proportional to g$ instead of p$ . Then we calculate only the 1PI diagrams. There exist three 1PI diagrams which contribute to GAL at one-loop level: (a) + loop, (b) + loop, (c) tad-pole, which are the same diagrams as those shown in Fig. 8 with replacing A$ with J5$ . [The Feynman rules for the propagators and the vertices in the Landau gauge are given in Appendix C.] These diagrams include ultraviolet divergences, which are renormalized by the parameters and 0elds. By taking a suitable subtraction scheme at zero temperature, all the divergences including 0nite corrections in the low energy limit are absorbed into the rede0nitions of the parameters and 0elds [103,95,96]. Then the loop diagrams generate only the temperature-dependent part. By using standard imaginary time formalism [140] we obtain

Nf a 5 1 (a) 2 4 GAL (p0 ; p I2 (T ) − J1 (M ; T ) + ˜ = 0) = {I4 (T ) − J1 (M ; T )} ; 2 2 6 3M2 (b) ˜ = 0) = GA L (p0 ; p

Nf a I2 (T ) ; 2 62

Nf 1 − a I2 (T ) ; (8.3) 2 2 where the functions In (T ) and Jmn (M ; T ) are de0ned in Appendix A.6. The total contribution is given by

Nf a 2 2 2 4 f (T ) = F − 2 I2 (T ) − aJ1 (M ; T ) + {I4 (T ) − J1 (M ; T )} : (8.4) 2 3M2 (c) ˜ = 0) = GA L (p0 ; p

67

Even when we use the transverse part instead of the longitudinal part to de0ne f (T ) in Eq. (8.2), we obtain the same result: GAT (p0 ; p ˜ = 0) = GAL (p0 ; p ˜ = 0).

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When we consider the low temperature region T M in the above expression, only the I2 (T ) term remains: f2 (T ) ≈ F2 −

Nf Nf 2 T : I2 (T ) = F2 − 2 2 12

(8.5)

which is consistent with the result given by Gasser–Leutwyler [82]. Thus, the pion decay constant decreases as T 2 dominated by the eAect of the thermal pseudoscalar mesons in the low temperature region. We should note that when we quantize only the 0eld, only the diagram (c) in Fig. 8 contributes and the resultant temperature dependence does not agree with the result by Gasser– Leutwyler. The above agreement is obtained from the fact that each diagram in Fig. 8 does generate the dominant contribution I2 (T ) = (2 =6)T 2 , and the terms proportional to aI2 (T ) are completely canceled among three diagrams. This cancellation is naturally understood as follows: The term pro(c) portional to aI2 (T ) in GA ˜ = 0) in Eq. (8.3) comes from the A–A–– vertex obtained from L (p0 ; p aLV term in the Lagrangian in Eq. (4.20), while the other term from the vertex from LA term. The vertices in the diagrams (a) and (b) in Fig. 8 come from aLV term. Then the above cancellation implies that the aLV term does not generate the thermal e/ect proportional to T 2 which is dominant in the low temperature region. The cancellation is similar to that occurred in the scattering: As was shown in Section 3.5, the aLV term generates the extra contact 4- interaction of order O(p2 ), which appears to violate the low energy theorem of the scattering amplitude. However, the aLV term also generates the -exchange contribution of order O(p2 ), which exactly cancels the contribution from the extra contact 4- interaction in the low energy region, Em . Thus, the aLV term does not generate the contribution of order O(p2 ). The similar cancellation occurs when the temperature is small enough compared with the meson mass, T m . As a result, the hadronic thermal e/ects is dominated by the contribution from LA term in the low temperature region, thus we obtained the result consistent with the “low temperature theorem” [82]. Let us estimate the critical temperature by naively extrapolating the above results to the higher temperature region. From Eq. (8.5) the critical temperature is well approximated as 12 (had) Tc ≈ F (0) ; (8.6) Nf where F (0) is the decay constant of at T = 0. In Ref. [102] the number of light >avors is chosen to be two, but here for later convenience we 0x Nf = 3. Then the critical temperature is given by Tc(had) ≈ 2F (0) 180 MeV :

(8.7)

It should be noticed that the above value of the critical temperature is changed only slightly even when we include the full e/ect given in Eq. (8.4) as shown in Ref. [102], as far as the vector meson mass is heavy enough: Tc(had) M . Next, let us study the corrections from the hadronic thermal e/ect to the mass. As was shown in Appendix C of Ref. [102], the and propagators are separated from each other in the Landau gauge, and the propagator takes simple form: − iD$ = −

PT$ PL$ − 2 : p2 − M2 + TVT p − M2 + TVL

(8.8)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

π

179

ρ

σ ρ

ρ π

σ (b)

(a)

ρ

ρ (c)

C

C (d)

(e)

Fig. 22. Feynman diagrams contributing to the vector meson self-energy in the Landau gauge: (a) loop, (b) loop, (c) loop, (d) tad-pole and (e) ghost-loop.

It is reasonable to de0ne the pole mass by using the longitudinal part in the low momentum limit, p ˜ = 0: 68 m2 (T ) = M2 − Re TVL (p0 = M ; p ˜ = 0; T ) ;

(8.9)

where Re TVL denotes the real part of TVL and inside the one-loop correction TVL we replaced m with M , since the di/erence is of higher order. The one-loop diagrams contributing to self-energy in the Landau gauge are shown in Fig. 22. Feynman rules for the vertices are shown in Appendix C. In Ref. [102], the divergences are renormalized in the on-shell renormalization scheme and the thermal corrections to the vector-meson two point function from the pseudoscalar and vector mesons are calculated. Thus, the parameter M in this section is renormalized in a way that it becomes the pole mass at T = 0. Since the calculation was done in the Landau gauge, the o/-shell structure of the propagator is not gauge invariant, while the result on mass–shell of vector meson is of course gauge invariant. Thus, here we show the thermal e/ects to the on-shell self-energy _T(p0 =M ; p ˜ ; T ) ≡ T(p0 =M ; p ˜ ; T )−T(p0 =M ; p ˜ ; T =0), in the low-momentum limit (˜ p = 0) by using the standard imaginary time formalism [140]: N f g 2 a2 G2 (M ; T ) ; ˜ = 0; T ) = Re _TVL(a) (p0 = M ; p 2 2 12 Nf g 2 1 G2 (M ; T ) ; ˜ = 0; T ) = Re _TVL(b) (p0 = M ; p 2 2 12 ˜ = 0; T ) Re _TVL(c) (p0 = M ; p Nf g 2 3 1 1 − F32 (M ; M ; T ) + F34 (M ; M ; T ) − F 6 (M ; M ; T ) = 2 2 2 2 3M2 3

4 1 G2 (M ; T ) ; − K6 (M ; M ; T ) + 3 12 68

It should be noticed that the transverse polarization agrees with the longitudinal one in the low momentum limit: TVT (p0 ; p ˜ = 0; T ) = TVL (p0 ; p ˜ = 0; T ).

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Re _TVL(d) (p0

Nf g 2 1 2 4 −2J1 (M ; T ) − = M ; p ˜ = 0; T ) = (I4 (T ) − J1 (M ; T )) ; 2 2 3M2

Nf g 2 1 G2 (M ; T ) ; (8.10) 2 2 6 where functions F, G, H , I , J and K are de0ned in Appendix A.6. Since the on-shell renormalization ˜ = 0; T = 0) = M2 , the sum of the above contributions is scheme implies that M2 + Re T(p0 = M ; p the thermal correction to the pole mass of meson. By noting that 1 1 1 n − F3n+2 (M ; M ; T ) = F3n (M ; M ; T ) − J (M ; T ) ; 2 3M 4 3M2 1 Re _TVL(e) (p0 = M ; p ˜ = 0; T ) = −

K6 (M ; M ; T ) = −

1 I4 (T ) ; 4M2

the thermal corrections to the vector meson pole mass is summarized as 69 2

Nf g 2 a 5 2 33 2 2 2 2 − G2 (M ; T ) + J1 (M ; T ) + M F3 (M ; M ; T ) : m (T ) = M + 22 12 4 16

(8.11)

(8.12)

Let us consider the low temperature region T M . The functions F and J are suppressed by e−M =T , and give negligible contributions. Noting that G2 (M ; T ) ≈ −(4 =15)T 4 =M2 for T M , the pole mass becomes m2 (T ) ≈ M2 −

N f g 2 a2 Nf 2 a 4 G2 (M ; T ) ≈ M2 + T : 2 2 12 360F2

(8.13)

Thus, the vector meson pole mass increases as T 4 at low temperature dominated by pion-loop eAect. The lack of T 2 -term is consistent with the result by the current algebra analysis [67]. 8.2. Vector manifestation at non-zero temperature In the analysis done in Ref. [102], the parameters F , a and g were assumed to have no temperature dependences, and the values at T = 0 were used. When we naively extrapolate the results in the previous subsection to the critical temperature, the resultant axialvector and vector current correlators do not agree with each other. Disagreement between the axialvector and vector current correlators is obviously inconsistent with the chiral symmetry restoration in QCD. However, the parameters of the HLS Lagrangian should be determined by the underlying QCD. As we explained in Section 5, the bare parameters of the (bare) HLS Lagrangian are determined by matching the HLS with the underlying QCD at the matching scale through the Wilsonian matching conditions. Since the quark condensate qq S as well as the gluonic condensate (#s =)G$ G $ in the right-hand side of the Wilsonian matching conditions (5.7)–(5.9) generally depends on the temperature, the application of the Wilsonian matching to the hot matter calculation implies that the bare parameters of the HLS (and hence M2 = ag2 F2 ) do depend on the temperature which are called the intrinsic temperature dependences [99] in contrast to the hadronic thermal e/ects. As is stressed in Ref. [99], 69

It should be stressed that this result is intact even when use the background 0eld gauge [100].

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181

the above disagreement is cured by including the intrinsic temperature dependences of the parameters through the Wilsonian matching conditions. In this subsection we brie>y review the analysis done in Ref. [99]. The intrinsic temperature dependences of the bare parameters lead to those of the on-shell parameters used in the analysis in the previous subsection through the Wilsonian RGEs. We write these intrinsic temperature dependences of the on-shell parameters explicitly as 70 F = F ($ = 0; T ) ; g = g($ = M (T ); T ) ; a = a($ = M (T ); T ) ;

(8.14)

where M is determined from the on-shell condition M = M (T ) = a($ = M (T ); T )g2 ($ = M (T ); T )F2 ($ = M (T ); T ) :

(8.15)

These intrinsic temperature dependences of the parameters give extra temperature dependences to the physical quantities which are not included by the hadronic thermal e/ects calculated in the previous subsection. Let us now apply the Wilsonian matching at the critical temperature Tc for Nf = 3. Here we assume qq S approaches to 0 continuously for T → Tc . 71 In such a case, the axialvector and vector current correlators derived from the OPE given in Eqs. (5.5) and (5.6) agree with each other. Then the Wilsonian matching requires that the axialvector and vector current correlators in the HLS given in Eqs. (5.2) and (5.3) must agree with each other. As we discussed in Section 6 for large Nf chiral restoration, this agreement is satis0ed if the following conditions are met [99]: g(; T ) → 0 ;

(8.16)

a(; T ) → 1 ;

(8.17)

T →Tc

T →Tc

z1 (; T ) − z2 (; T ) → 0 : T →Tc

(8.18)

The conditions for the parameters at the matching scale g(; Tc )=0 and a(; Tc )=1 are converted into the conditions for the on-shell parameters through the Wilsonian RGEs in Eqs. (4.211) and (4.210). Since g = 0 and a = 1 are separately the 0xed points of the RGEs for g and a, the on-shell parameters also satisfy (g; a) = (0; 1), and thus M = 0. Noting that G2 (M ; T ) → I2 (T ) =

2 2 T ; 6

J12 (M ; T ) → I2 (T ) =

2 2 T ; 6

M →0

M →0

70 We note that $ in Eq. (8.14) is the renormalization scale, not a chemical potential. In the next subsection where we for expressing the chemical potential. consider the dense matter calculation, we use $ 71 It is known that there is no Ginzburg–Landau type phase transition for Nf = 3 (see, e.g., Refs. [194,43]). There may still be a possibility of non-Ginzburg–Landau type continuous phase transition such as the conformal phase transition [148].

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M2 F32 (M ; M ; T ) → 0 ; M →0

(8.19)

Eq. (8.12) in the limit M T reduces to Nf 15 − a2 I2 (T ) : (8.20) 22 12 Since a 1 near the restoration point, the second term is positive. Then the pole mass m is bigger than the parameter M due to the hadronic thermal corrections. Nevertheless, the intrinsic temperature dependence determined by the Wilsonian matching requires that the vector meson becomes massless at the critical temperature: m2 (T ) = M2 + g2

m2 (T ) → 0 ; T →Tc

(8.21)

since the 0rst term vanishes as M → 0, and the second term also vanishes since g → 0 for T → Tc . This implies that, as was suggested in Refs. [106,44,45], the vector manifestation (VM ) actually occurs at the critical temperature [99]. This is consistent with the picture shown in Refs. [42–45]. We should stress here that the above m (T ) is the pole mass of meson, which is important for analysing the dilepton spectra in RHIC experiment. It is noted [106] that although conditions for g(; T ) and a(; T ) in Eqs. (8.16) and (8.17) coincide with the Georgi’s vector limit [85,86], the VM here should be distinguished from Georgi’s vector realization [85,86]. Let us determine the critical temperature. For T ¿ 0 the thermal averages of the Lorentz non-scalar operators such as q S $ D q exist in the current correlators in the OPE [108]. Since these contributions are small compared with the main term 1 + #s =, we expect that they give only small corrections to the value of the critical temperature, and neglect them here. Then, the Wilsonian matching condition to determine the bare parameter F (; Tc ) is obtained from that in Eq. (5.8) by taking qq S = 0 and including a possible temperature dependence of the gluonic condensate:

#s 22 (#s =)G$ G $ Tc 1 F2 (; Tc ) 1 + ; (8.22) = + 2 82 3 4 which determines the on-shell parameter F ($ = 0; Tc ) through the Wilsonian RGE for F in Eq. (4.208) with taking (g; a) = (0; 1). It should be noticed that the F ($; T ) does run with scale $ by the Wilsonian RGE [104,105] even at the critical point. As we obtained for large Nf in Eq. (6.106), the relation between F (; Tc ) and F ($ = 0; Tc ) is given by Nf F2 (0; Tc ) F2 (; Tc ) = − : 2 2(4)2 2

(8.23)

On the other hand, the relation between F (0; Tc ) and the physical pion decay constant, which of course vanishes at T = Tc , is given by taking M = 0 and a = 1 in Eq. (8.4) [99]: Nf Nf 2 0 = f2 (Tc ) = F2 (0; Tc ) − 2 I2 (Tc ) = F2 (0; Tc ) − (8.24) T : 4 24 c Here we should note that the coeVcient of I2 (Tc ) in the second term is a half of that in Eq. (8.5) which is an approximate form for T M taken with assuming that the vector meson does not become light. The factor 1=2 appears from the contribution of which becomes the real NG boson at the critical temperature due to the VM. This situation is similar to that occurring in the coeVcients

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Table 19 Estimated values of the critical temperature Tc for several choices of the value of the matching scale with QCD = 400 MeV Tc

0:8 0:21

0:9 0:22

1:0 0:23

1:1 0:25

Units of and Tc are GeV.

of the quadratic divergences in the solution of the RGE for F : In Eq. (4.219) only the quadratic divergence from the pion loop is included, while in Eq. (6.106) that from the loop (-loop) is also included. Then the extra factor 1=2 appears in the second term of Eq. (6.106) compared with that of Eq. (4.219). From Eq. (8.24) together with Eqs. (8.22) and (8.23) the critical temperature is given by

1=2 $ Nf #s 22 (#s =)G$ G Tc 24 32 Tc = 1+ + F (0; Tc ) = − : (8.25) Nf Nf 2 3 4 4 Let us estimate the critical temperature for Nf =3. The value of the gluonic condensate near phase transition point becomes about half of that at T = 0 [144,45], so we use "# # s G$ G $ = 0:006 GeV4 ; (8.26) obtained by multiplying the value at T = 0 shown in Refs. [171,172] [see Eq. (5.13)] by 1=2. For the value of the QCD scale QCD we use QCD = 400 MeV ;

(8.27)

as a typical example. For this value of QCD , as we showed in Tables 9 and 10 in Section 5.3, the choice of 1:1 GeV for the values of the matching scale provides the predictions in good agreement with experiment at T = 0. However, the matching scale may have the temperature dependence. In the present analysis we use = 0:8; 0:9; 1:0 and 1:1 GeV ;

(8.28)

and determine the value of the critical temperature Tc from Eq. (8.25). We show the resultant values in Table 19. We note that the estimated values of Tc in Table 19 are larger than that in Eq. (8.7) which is obtained by naively extrapolating the temperature dependence from the hadronic thermal e/ects without including the intrinsic temperature dependences. This is because the extra factor 1=2 appears in the second term in Eq. (8.24) compared with that in Eq. (8.5). As we stressed below Eq. (8.24), the factor 1=2 comes from the contribution of (longitudinal ) which becomes massless at the chiral restoration point. The vector dominance in hot matter and the dependences of the critical temperature on other parameter choices will be studied in Ref. [100].

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8.3. Application to dense matter calculation In this section we brie>y review the application of the Wilsonian matching and the vector manifestation (VM) to the dense matter calculation done in Ref. [93]. To set up the arguments for the density problem, we consider a system of hadrons in the background of a 0lled Fermi sea. For the moment, we consider the Fermi sea as merely a background, side-stepping the question of how the Fermi sea is formed from a theory de0ned in a matter-free vacuum. Imagine that mesons—the pion and the meson—are introduced in HLS with a cuto/ set at the scale , . Since we are dealing with dense fermionic matter, we may need to introduce the degrees of freedom associated with baryons or alternatively constituent quarks (or quasiquarks) into the HLS. At low density, say, n ¡ n, ˜ with n˜ being some density greater than n0 , the precise value of which cannot be pinned down at present, we may choose the cuto/ 0 below the nucleon mass, mN ∼ 1 GeV, but above the mass m = 770 MeV and integrate out all the baryons. In this case, the bare parameters of the HLS Lagrangian will depend upon the density n (or equivalently Fermi momentum PF ) since the baryons that are integrated out carry information about the baryon density through their interactions in the full theory with the baryons within the Fermi sea. Once the baryons are integrated out, we will then be left with the standard HLS Lagrangian theory with the NG and gauge boson 0elds only except that the bare parameters of the eAective Lagrangian will be density-dependent. It should be noticed that the cutoA can also be density dependent. However, in general, the density-dependence of the cuto/ is not related to those of the bare parameters by the RGEs. For T ¿ 0 and n = 0 this di/erence appears from the “intrinsic” temperature dependence introduced in Ref. [99] (see previous subsection) which was essential for the VM to occur at the chiral restoration point. As density increases beyond n, ˜ the fermions may however start 0guring explicitly, that is, the fermion 0eld may be present below the cuto/ ˜ (n ¿ n). ˜ The reason is that as density approaches the chiral restoration point, the constituent-quark (called quasiquark) picture—which seems to be viable even in matter-free space [176]—becomes more appropriate [45] and the quasiquark mass drops rapidly, ultimately vanishing (in the chiral limit) at the critical point. This picture has been advocated by several authors in a related context [164]. To study the e/ects of the quasiquark near the critical density in Ref. [93] the HLS with the quasiquark was adopted. There a systematic counting scheme was introduced into the model and a systematic derivative expansion similar to the one explained in Section 4 was made. In the HLS with the quasiquark (constituent quark) the quasiquark 0eld is introduced in the Lagrangian in such a way that it transforms homogeneously under the HLS: → h(x) · where h(x) ∈ Hlocal . Since we consider the model near chiral phase transition point where the quasiquark mass is expected to become small, we assign O(p) to the quasiquark mass mq . Furthermore, we assign O(p) to the 72 or the Fermi momentum PF , since we consider that the cuto/ is larger than chemical potential $ even near the phase transition point. Using this counting scheme we can make the systematic $ expansion in the HLS with the quasiquark included. We should note that this counting scheme is di/erent from the one in the model for and baryons given in Ref. [142] where the baryon

72

In Ref. [93] $ is used for expressing the chemical potential. Throughout this report, however, we use $ for the energy for the chemical potential in this subsection. scale, and then we use $

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mass is counted as O(1). The leading order Lagrangian including one quasiquark 0eld and one anti-quasiquark 0eld is counted as O(p) and given by [24,93] 0 − mq ) (x) LQ(1) = S (x)(iD$ $ − $ + S (x)(K $ #ˆ $ (x) + 6 5 $ #ˆ⊥$ (x)) (x) ;

(8.29)

where D$ = (9$ − ig$ ) and K and 6 are constants to be speci0ed later. At one-loop level Lagrangian (8.29) generates the O(p4 ) contributions including hadronic denseloop e/ects as well as divergent e/ects. The divergent contributions are renormalized by the parameters, and thus the RGEs for three leading order parameters F , a and g (and parameters of O(p4 ) Lagrangian) are modi0ed from those without quasiquark 0eld. In addition, we need to consider the renormalization group >ow for the quasiquark mass mq . 73 Calculating one-loop contributions for RGEs, we 0nd [93] m2q 2 Nf dF2 2 2 2 2 $ = [3a g F + 2(2 − a)$ ] − 2 6 Nc ; d$ 2(4)2 2

Nf 62 m2q $2 da 2 =− + a (a − 1) 3a(1 + a)g − (3a − 1) Nc ; $ d$ 2(4)2 F2 22 F2

(8.30) (8.31)

$

Nf 87 − a2 4 Nc dg2 =− g + 2 g4 (1 − K)2 ; 2 d$ 2(4) 6 6

(8.32)

$

mq dmq = − 2 [(C − C )$2 − m2q (C − C ) + M2 C − 4C ] ; d$ 8

(8.33)

where

C ≡ C ≡

6 F K F

2 2

Nf2 − 1 ; 2Nf Nf2 − 1 ; 2Nf

C ≡ g2 (1 − K)2

Nf2 − 1 : 2Nf

(8.34)

Hadronic dense corrections from the quasiquark loop to the decay constant f ( $) and the pole mass m ( $) were calculated in Ref. [93]. Here we will brie>y review the analysis. As for the calculation of the hadronic thermal corrections explained in the previous subsections, it is convenient to use the following “on-shell” quantities: ) ; F = F ($ = 0; $ ); g = g($ = M ( $); $ 73

) ; a = a($ = M ( $); $

(8.35)

=$ c , K = 6 = 1 while at $ ¡$ c , K = 6. The running will be small The constants K and 6 will also run such that at $ near nc , so we will ignore their running here.

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where M is determined from the “on-shell condition”: ) $) = a($ = M ( $); $ M2 = M2 ( )F2 ($ = M ( ) : $); $ $); $ ×g2 ($ = M (

(8.36)

Then, as in the previous subsection, the parameter M in this subsection is renormalized in such a = 0. way that it becomes the pole mass at $ For obtaining the dense-loop corrections to the pion decay constant we should note that distinction has to be made between the temporal and spatial components of the pion decay constants, since the Lorentz invariance is broken in the medium. We use the following de0nition [161]: 0|J5$=0 (0)|(˜ p) $ = −ip0 ft ( $) ; 0|J5$=i (0)|(˜ p) $ = −i˜ pi fs ( $) :

(8.37) $

In terms of the axialvector–axialvector two-point function TAA , the temporal and spatial components of the pion decay constant are generally expressed as $ u T (p ; p ˜ )p $ 0 1 AA ft ( $) = ; p0 F p0 =!˜

$ # −p (g − u u )T (p ; p ˜ )p #$ # $ 0 1 AA $) = fs ( 2 pS F

;

(8.38)

p0 =!˜

is the wave function renormalization constant in medium. 74 According to the analysis where F is nothing but ft : of Ref. [142] in dense matter, this F = ft ( F $) : (8.39) $ is expressed as In the HLS with present renormalization scheme, this TAA $ $ $ 2 $ 2 $ ˜ ) = g F + 2z2 (g p − p p ) + TS AA (p0 ; p ˜) ; TAA (p0 ; p

(8.40)

(p ; p ˜ ) denotes the hadronic dense corrections of interest. In Ref. [93] the dense-loop where TS $ AA 0 corrections from the interaction Lagrangian (8.29) were calculated at one loop, and it was shown that there is no hadronic dense-loop correction to the decay constants: ) : [ft ( $)]2 = ft ( $)fs ( $) = F2 ($ = 0; $

(8.41)

Next we calculate the hadronic dense-loop corrections to the pole mass. As in the previous subsection there are two pole masses related to the longitudinal and transverse components of propagator. In Ref. [93] the dense-loop corrections to them from the Lagrangian (8.29) were calculated at one-loop level. The results are $) = m2T ( $) m2L ( = M2 +

2 2 g (1 − K)2 [BS S − (M2 + 2m2q )BS 0 (p0 = M ; p ˜ = 0)] ; 3

(8.42)

+ · · ·. For $ = 0 this F Note that the background 0eld A includes the background pion 0eld S as A$ = A$ + 9$ = SF agrees with F . 74

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where

1 P + ! F F 2 ; BS S = 2 PF !F − mq ln 4 mq PF + !F 1 S ˜ = 0) = 2 −ln B0 (p0 ; p 8 mq  2 2 − p2 − ij + P 2 2 4m −p − ij ! F F q 0 0 1 4mq − p0 − ij  ; + (8.43) ln 2 2 −p0 − ij 2 2 2 !F 4mq − p0 − ij − PF −p0 − ij with PF being the Fermi momentum of quasiquark and !F = PF2 + m2q . c for Nf = 3. Here we Let us now apply the Wilsonian matching at the critical chemical potential $ note that the current correlators in the HLS remains unchanged as the forms given in Eqs. (5.2) and (5.3) except that the bare parameters are density-dependent even when we include the quasiquark 0eld as explained above. 75 As is done for the chiral restoration in hot matter in Ref. [99] (see previous subsection) we assume that qq S approaches to 0 continuously for n → nc . 76 In such a case the axialvector and vector current correlators by OPE in the QCD sector approach each other, and will agree at nc . Then, through the Wilsonian matching we require that the correlators in he HLS in Eqs. (5.2) and (5.3) agree with each other. As in the case of large Nf [106] (see Section 6) and in the case of T ∼ Tc [99] (see Section 8.2), this agreement can be satis0ed also in dense matter if the following conditions are met [93]: g(; n) → 0; n→nc

a(; n) → 1 ; n→nc

z1 (; n) − z2 (; n) → 0 : n→nc

(8.44)

The above conditions for the bare parameters are converted to the ones for the on-shell parameters through the Wilsonian RGEs given in Eqs. (8.30)–(8.33). Di/erently from the cases for large Nf QCD and hot QCD, Eqs. (8.31) and (8.32) show that (g; a) = (0; 1) is a 0xed point only when mq = 0. Since the “on-shell” quasiquark mass mq is expected to vanish at the critical point: mq (n) → 0 ; n→nc

(8.45)

and that mq =0 is actually a 0xed point of the RGE in Eq. (8.33), (g; a; mq )=(0; 1; 0) is a 0xed point of the coupled RGEs for g, a and mq . Furthermore and most importantly, X = 1 becomes the ?xed point of the RGE for X [107]. This means that at the 0xed point, F (0) = 0 [see Eq. (4.254)]. What = 0, this F (0) = 0 does this mean in dense matter? To see what this means, we note that for T = $ condition is satis0ed for a given number of >avors Nfcr ∼ 5 through the Wilsonian matching [106]. = 0 and T = 0, this condition is never satis0ed due to thermal hadronic corrections For Nf = 3, $ 75 Since the Lorentz non-invariant terms in the current correlators by the OPE are suppressed by some powers of n=3 (see, e.g. Ref. [110]), we ignore them from both the hadronic and QCD sectors. 76 We are assuming that the transition is not strongly 0rst order. If it is strongly 0rst order, some of the arguments used here may need quali0cations. However, we should note that, in the presence of the small current quark mass, the quark condensate is shown to decrease rapidly but continuously around the “phase transition” point [43].

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[99]. Remarkably, as was shown in Ref. [93] and we brie>y reviewed above, for Nf = 3, T = 0 and $ = $c , it turns out that dense hadronic corrections to the pion decay constant vanish up to O(p6 ) corrections. Therefore the 0xed point X = 1 (i.e., F (0) = 0) does indeed signal chiral restoration at the critical density. c . As is shown Let us here focus on what happens to hadrons at and very near the critical point $ in Eq. (8.41), there is no hadronic dense-loop corrections to the decay constants. Thus c ) = 0 : f ( $c ) = F (0; $

(8.46)

Since c ) = F2 (; $ c ) − F2 (0; $

Nf 2 ; 2(4)2

(8.47)

c ) is given by a QCD correlator at $ =$ c , $ c can be computed and at the matching scale , F2 (; $ from Nf c ) = F2 (; $ 2 : (8.48) 2(4)2 Note that in free space, this is the equation that determines Nfcrit ∼ 5 [106]. In order for this equation c )=F2 (; 0) ∼ 3=5. We do not to have a solution at the critical density, it is necessary that F2 (; $ have at present a reliable estimate of the density dependence of the QCD correlator to verify this condition but the decrease of F of this order in medium looks quite reasonable. c . For M ; mq PF Eq. (8.42) reduces to Next we compute the pole mass near $ 2 $ m2 ( $) = M2 ( $) + g2 2 (1 − K)2 : (8.49) 6 =$ c , we have g = 0 and a = 1 so that M ( At $ $) = 0, and then m ( $) = 0. Thus the fate of the meson at the critical density is the same as that at the critical temperature [93]: m2 ( $) → 0 : $→ $c

(8.50)

This implies that, as was suggested in Ref. [106] and then proposed in Refs. [44,45], the vector manifestation (VM) is realized in dense matter at the chiral restoration with the mass m going to = 0, zero at the critical point. Thus the VM is universal in the sense that it occurs at Nfcrit for T = $ = 0 and at $ c for T = 0 and Nf ¡ Nfcrit . at Tc for Nf ¡ Nfcrit and $ Detailed calculations of the hadronic dense-loop corrections are shown in Ref. [93], where the O(p2 ) interaction Lagrangian was included in addition to the Lagrangian in Eq. (8.29) and it was shown that the results in Eqs. (8.46) and (8.50) are intact. 9. Summary and discussions In this report we have explained recent development, particularly the loop e/ects, of the e/ective 0eld theory (EFT) of QCD and QCD-like theories for light pseudoscalar and vector mesons, based on the hidden local symmetry (HLS) model. The HLS model as explained in Section 3 is simply reduced to the nonlinear chiral Lagrangian in the low energy region where the kinetic term of the vector meson is negligible compared with the

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mass term, p2 m2 , and the gauge symmetry (gauge-boson degree of freedom) becomes “hidden”. Although there are many vector meson theories which yield the same classical (tree level) result as that of the HLS model, they may not lead to the same quantum theory. Actually, as was illustrated in the CP N −1 model [24], theories being the same at classical (tree) level, the one with explicit gauge symmetry and the other without it, may not be the same at quantum level. In Section 4 it was emphasized that presence of the gauge symmetry of HLS is in fact vital to the systematic low-energy expansion (chiral perturbation) with loops of vector as well as pseudoscalar mesons, when their masses can be regarded as small. We developed a systematic expansion of HLS model on the same footing as the chiral perturbation theory (ChPT) of the ordinary chiral Lagrangian without vector mesons (reviewed in Section 2), based on the order counting of the HLS coupling g: g ∼ O(p) :

(9.1)

Based on this systematic expansion, we developed in Section 4 analyses of the one-loop renormalization-group equations (RGEs) in the sense of Wilson (“Wilsonian RGE”) which includes quadratic divergence. The Lagrangian having such running parameters corresponds to the Wilsonian e/ective action which is obtained from the bare action (de0ned at cuto/) by integrating higher energy modes down to lower energy scale and necessarily contains quadratic divergences. Here we should emphasize that, as a matter of principle, the bare parameters of EFT are not free parameters but are determined by the underlying theory and hence the quadratic divergences should not be renormalized out by cancelling with the arbitrary choice of the bare parameters as in the usual renormalization procedure. Once we determined the bare parameters of EFT, we necessarily predict the physical quantities through the Wilsonian RGEs including the quadratic divergence. A novel feature of the approach in this report is the “Wilsonian matching” given in Section 5, which determines the bare parameters (de0ned at the cuto/ scale ) of the EFT in terms of the underlying theory, the QCD or QCD-like theories. We wrote down the current correlators at in terms of the bare parameters of the HLS model, which was then evaluated in terms of the OPE of the underlying QCD at the same scale . This provides the EFT with otherwise unknown information of the underlying theory such as the explicit dependence on Nc and QCD as well as the precise value of the bare parameters. Once the bare values were given as the boundary conditions of RGEs, the physics below was uniquely predicted via RGEs through the own dynamics of the HLS model. Main issues of this approach were: 1. Prediction of a very successful phenomenology of and for the realistic case of Nf = 3 (Section 5). 2. Prediction of chiral symmetry restoration due to quadratic divergence for certain choice of the parameters of the underlying QCD, such as the number of colors Nc and of the massless >avors Nf such that Nf =Nc ¿ 5=3 (Section 6). The vector meson dominance, though accidentally valid for the Nf = 3, does not hold in general and is largely violated near the chiral restoration point. 3. Prediction of “vector manifestation (VM)” as a novel feature of this chiral restoration: The becomes the chiral partner of the in contrast to the conventional manifestation of the linear sigma model (“GL manifestation”) where the scalar meson becomes the chiral partner of the (Section 6). Similar phenomenon can also take place in the hot/dense QCD (Section 8). 4. The chiral restoration in the HLS model takes place by its own dynamics as in the underlying QCD, which suggested that the Seiberg-type duality is operative even for the non-SUSY QCD

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where the HLS plays a role of the “magnetic gauge theory” dual to the QCD as the “electric gauge theory” (Section 6). It was demonstrated in Section 4 that the quadratic divergences are actually vital to the chiral symmetry restoration in the EFT which corresponds to the underlying QCD and QCD-like theories under extreme conditions where such a chiral phase transition is expected to take place. The point is that the quadratic divergence in the HLS model gives rise to an essential part of the running of the decay constant F2 ($) whose bare value F2 () is not the order parameter but merely a Lagrangian parameter, while the pole residue of the NG boson is proportional to the value F2 (0) at the pole position p2 = 0 which is then the order parameter of the chiral symmetry breaking. The chiral restoration is thus identi0ed with F2 (0) = 0, while F2 () = 0 in general. We gave detailed explanation why the quadratic/power divergence is so vital to the phase transition of the EFT, based on the illustration of the phase transitions in various well-known models having the chiral phase transition, like the NJL model, the standard model (SM) and the CP N −1 in D(6 4) dimensions as well as the nonlinear chiral Lagrangian which is of direct relevance to our case. The point is that the bare Lagrangian as it stands does not tell us which phase we are actually living in. The quadratic divergence is the main driving force to make the quantum theory to choose a di/erent phase than that the bare Lagrangian looks like. Now, one might suspect that the systematic expansion in our case might break down when we include the quadratic divergence. Actually the quadratic divergence carrying no momentum would not be suppressed by powers of p in the HLS model as well as in the ChPT (with the quadratic divergence included): Quadratic divergences from all higher loops would in principle contribute to the O(p2 ) term in powers of [Nf 2 =(4F ())2 ]n for nth loop and hence would invalidate the power counting rule in the systematic expansion unless Nf 2 =(4F ())2 ¡ 1. However, such a condition is needed even in the usual ChPT (without quadratic divergence, F () = F (0) ≡ F ) where the systematic expansion breaks down unless Nf 2 =(4F )2 ¡ 1. Inclusion of the quadratic divergence is actually even better for the systematic expansion to work, Nf

2 ¡1 ; (4F ())2

(9.2)

since generally we have F2 () ¿ F2 (0) due to quadratic divergence, and in particular near the chiral restoration point where F2 (0) → 0 whereas F2 () remains 0nite. More speci0cally, F2 () was given by the Wilsonian matching with the QCD (Eq. (5.21)): 2 2 Nc 2 ∼ Nc ; (9.3) F () = 2(1 + A ) 3 4 4 where A stands for the higher order corrections in OPE to the parton (free quark loop) contribution 1 and hence is expected to be A 1. Actually we estimated A ∼ 0:5 for Nf =3. Then the systematic expansion would be valid if Nf

Nf 2 ∼ ¡1 : (4F ())2 Nc

(9.4)

Such a situation can be realized, if we consider the large Nc limit Nc → ∞ such that Nf =Nc 1 and then extrapolate it to the parameter region Nf =Nc ∼ 1. Moreover, in the HLS model

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(in contrast to ChPT without vector meson), the quadratic divergence for F2 has an additional factor 1=2 (at a 1) and hence the systematic expansion is expected to be valid for Nf Nf 2 ∼ ¡1 : 2 2 (4F ()) 2Nc

(9.5)

Thus the inclusion of the quadratic divergence does not a/ect the validity of the systematic expansion. It even improves the scale for the systematic expansion better than the conventional naive dimensional analysis (without quadratic divergence). Note that the edge of the validity region of the systematic expansion roughly corresponds to the chiral restoration point where the tree and the loop cancel out each other. Actually, the phase transition in many cases is a phenomenon in which the tree (bare) and the loop e/ects (quadratic divergences) are becoming comparable and are balanced (cancelled) by each other. Hence this phenomenon is generally at the edge of the validity of the systematic expansion, such as in the usual perturbation (SM), chiral perturbation (nonlinear sigma model), etc., although in the NJL case the loop to be balanced by the tree is treated also as the leading order in the 1=N expansion. To summarize the roles of the quadratic divergence: It must be included as a matter of principle once the bare parameters are 0xed; It is crucial to the phase transition; It improves the validity scale of the systematic expansion rather than naive dimensional analysis; It leads to a very successful phenomenology of and system. Now, once we matched the EFT, the HLS model, with the underlying theory in this way, we can play with arbitrary Nc and QCD as well as Nf in the same sense as dealing with the underlying QCD. Then we expect that the HLS model by its own dynamics will give rise to the same infrared physics as the underlying theory itself for arbitrary parameter choice other than Nc = Nf = 3 of the real life QCD: When the underlying QCD gets chiral restoration, the HLS model will also get chiral restoration. In Section 6 we actually formulated conditions of chiral symmetry restoration on the bare HLS parameters (“VM conditions”) by matching the current correlators with those of the underlying QCD where the chiral symmetry gets restored, qq S → 0 as Nf → Nfcrit : g() → 0; a() → 1; z1 () − z2 () → 0 ; 2 Nc 2 crit 2 crit 2(1 + A ) F () → (F ) ≡ ; 3 4

(9.6)

where crit S =0 = A ≡ A |qq

3(Nc2 − 1) #s 22 (#s =)G$ G $ + 8Nc Nc 4

(9.7)

must satisfy 0 ¡ crit A ¡ 1 in order that the OPE makes sense. Although the VM conditions as they stand might not seem to indicate chiral symmetry restoration, they actually lead to the vanishing order parameter F2 (0) → 0 and thus the chiral restoration through the own dynamics of the HLS model as follows: The RGEs of the HLS model are readily solved for the VM conditions, since g = 0 and a = 1 are the 0xed points of the RGEs.

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By taking g → 0 and a → 1, we had 2 crit N 2 1 N f c 2(1 + crit F2 (0) = F2 () − Nf → ) − ; A 2 4 3 2 4

(9.8)

where the 0rst term given by the Wilsonian matching with QCD is proportional to Nc and the second term given by the quadratic divergence in the HLS model is proportional to Nf . Now, the chiral restoration takes place with vanishing right-hand side (RHS), F2 (0) → 0, by precise cancellation between the two terms, namely the interplay between Nc and Nf such that Nf ∼ Nc 1. Then the chiral restoration takes place at Nc Nf = Nfcrit = 4(1 + crit (9.9) A ) ; 3 crit where 0 ¡ crit A ¡ 1 in order for the OPE to make sense. Then we predicted the critical value Nf fairly independently of the detailed input data: Nc Nc crit ¡ Nf ¡ 8 ; (9.10) 4 3 3 which is consistent with the lattice simulation [118] 6 ¡ Nfcrit ¡ 7

(Nc = 3) ;

(9.11)

but in disagreement with the analysis of ladder Schwinger–Dyson (SD) equation combined with the perturbative infrared 0xed point [12,14]; Nc : (9.12) Nfcrit 12 3 More speci0cally, we estimated crit S = 0 (A ∼ 0:5 A 0:25 at the QCD chiral restoration point qq for Nc = Nf = 3 where qq S = 0) and hence Nc crit ; (9.13) Nf 5 3 which coincides with the instanton argument [182]. It was emphasized that this chiral restoration should be regarded as a limit F2 (0) → 0 but not precisely on the critical point F2 (0) ≡ 0 where no light composite spectrum would exist and hence the HLS model would break down. The limit (“VM limit”) corresponds to the VM conditions for the bare parameters; F2 () → (Fcrit )2 , a() → 1 and g() → 0 as Nf → Nfcrit in the underlying QCD, with a special care for the g() → 0, in contrast to setting g() ≡ 0 which gives the HLS model a redundant global symmetry, G1 ×G2 with G =SU(Nf )L ×SU(Nf )R , larger than that of the underlying QCD and should be avoided. On the other hand, there is no peculiarity for setting a() = 1 as far as we keep g() = 0, in which case the redundant global symmetry G1 × G2 is explicitly broken only by the gauge coupling down to the symmetry of the HLS model, Gglobal × Hlocal . In the real-life QCD with Nf = 3 which we showed is very close to a() = 1, this coupling is rather strong. It is amazing, however, that by simply setting Nf → Nfcrit in the underlying QCD, we arrive at the VM limit which does realize the weak coupling gauge theory of light composite , g → 0 and m → 0, in spite of the fact that this coupling is dynamically generated at composite level from the underlying strong coupling gauge theory.

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The salient feature of the above chiral restoration is that the becomes the chiral partner of the with its mass vanishing at the critical point: m2 → m2 = 0;

F2 (m )=F2 (0) → 1 ;

(9.14)

as F2 (0) → 0, where F (m ) is the decay constant of (longitudinal ) at on-shell. This we called “vector manifestation (VM)” in contrast to the conventional manifestation aU la linear sigma model (“Ginzburg–Landau/Gell-Mann–Levy(GL) manifestation”): m2S → m2 = 0 ; F2 (0)

(9.15)

as → 0, where mS stands for the mass of the scalar meson (“sigma” meson in the linear sigma model). The VM implies that belongs to (Nf2 − 1; 1) ⊕ (1; Nf2 − 1) of the chiral representation together with the , while in the GL manifestation does to (Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf ) together with the scalar meson. The GL manifestation does not satisfy the Wilsonian matching: since the vector current correlator has no scalar meson contributions, we would have TV = 0, were it not for the contribution, and hence −Q2 (d=dQ2 )TV |Q2 =2 = 0, whereas QCD yields non-zero value −Q2 (d=dQ2 )TV(QCD) |Q2 =2 = 2 (1+crit V )Nc =(24 ) = 0. The vanishing TV together with the restoration requirement TA =TV would imply TA = 0, which is in contradiction with the Wilsonian matching for TA : Q2 (d=dQ2 )TA |Q2 =2 = 2 F2 ()=2 = −Q2 (d=dQ2 )TA(QCD) |Q2 =2 = (1 + crit A )Nc =(24 ) = 0. The fact that both the e/ective theory and the underlying theory give the same infrared physics is an aspect of the duality of Seiberg-type 0rst observed in the SUSY QCD: in the case at hand, non-SUSY QCD, we found that the HLS plays a role of the “magnetic gauge theory” dual to the QCD as the “electric gauge theory”. Here we recall that the phase structure of the SUSY QCD was revealed by Seiberg only in terms of the e/ective theory in the sense of Wilsonian e/ective action. In this paper we have demonstrated that the same is true also in the non-SUSY QCD, namely the Wilsonian RGEs (including the quadratic divergence) in the e/ective 0eld theory approach are very powerful tool to investigate the phase structure of the QCD and the QCD-like gauge theories. In Section 7, we gave a brief review of the proof of the low-energy theorem of the HLS, g = 2g F2 at any loop order following Refs. [95,96]. We showed that the inclusion of the quadratic divergence does not change the proof, which implies that the low-energy theorem of the HLS is valid at any loop order even under the existence of the quadratic divergence. Finally in Section 8, we gave a brief review on the application of the approach explained in previous sections to the hot and/or dense matter calculation based on Refs. [99,93]. We have summarized how the VM takes place at the chiral restoration point in hot matter at zero density [99] and also in dense matter at zero temperature [93]. The picture based on the VM in hot matter would provide several peculiar predictions on, e.g., the vector and axialvector susceptibilities [94], the vector dominance of the electromagnetic form factor of pion [100], and so on which can be checked in the experiments in operation as well as in future experiments. These analysis are still developing, so we did not include the review in this report. We encourage those who have interest to read Refs. [99,93,94,100]. Several comments are in order: One might suspect that the limit of m → 0 would be problematic since the on-shell amplitude would have a factor 1=m2 in the (longitudinal) polarization tensor j(0) $ and thus divergent in such a limit. In our case such a polarization factor is always accompanied by a gauge coupling g2 of

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, which yields the amplitude a factor g2 =m2 ∼ 1=F2 (m ) ∼ 1=F2 (m ). Then the above problem is not a peculiarity of the massive vector mesons in our HLS model but is simply reduced to the similar problem as in the nonlinear sigma model at the chiral restoration point. In the nonlinear chiral Lagrangian without quadratic divergence, the on-shell amplitude like – scattering behaving like A(p2 ) = p2 =F2 (0) [see Eq. (3.83)] would be divergent at the chiral restoration point F (0) → 0, which simply implies that the EFT is valid only for p ¡ F (0), namely the validity region is squeezed out at the restoration point F (0) → 0. Such a problem does not exist in the linear sigma model (or Higgs Lagrangian in the SM) thanks to the light scalar meson (Higgs boson) introduced in addition to the NG boson . 77 However, in our case where the quadratic divergence is included, the problem is also solved in a similar way even without the additional scalar meson as follows: The amplitude is expected to behave as p2 =F2 (p2 ) ∼ X ($2 = p2 ), with X ($) de0ned in Eq. (4.254), which is non-singular in the m → 0 limit as we have discussed around the end of Section 6.1.5. Actually, the amplitude A(s) ∼ s=F2 (s) = X (s) has a vanishing low-energy limit X (0) = 0, as far as we approach the chiral restoration point F2 (0) → 0 from the broken phase F2 (0) = 0, i.e., the VM limit with m → 0. On the other hand, when the theory is exactly on the VM point, X (s) is a certain (non-zero) constant, i.e., X (s) ≡ (constant) which leads to X (s) → (constant) = 0 even at the s → 0 limit. Although the low-energy limit amplitude A(0) is discontinuous across the phase transition point, the amplitude is non-singular at the phase transition point similarly to the linear sigma model, in sharp contrast to the conventional nonlinear sigma model without quadratic divergence. Thus our case is a counter example against the folklore that the massive vector meson theory has a problem in the massless limit, unless the mass is via Higgs mechanism with the additional light scalar meson (Higgs boson). The axialvector mesons A1 including a1 are heavier than the matching scale, = 1:1–1:2 GeV, so that we did not include them in the analysis based on the Wilsonian matching in Section 5. It was checked that even including A1 does not substantially change the value of F2 () given by the OPE and hence does not a/ect the qualitative feature of our analysis for Nf = 3. We then expect that the A1 in the VM is resolved and/or decoupled from the axialvector current near the critical point, since the is already balanced with the and there is no contribution in the vector current correlator to be matched with the additional contribution in the axialvector current correlator. On the other hand, the recent analyses [97,98,181,115,149,124] show that there exist light scalar mesons, some of which has a mass smaller than our matching scale 1:1 GeV. However, the scalar meson does not couple to the axialvector and vector currents, anyway. We expect that the scalar meson is also resolved and/or decoupled near the chiral phase transition point, since it is in the pure (Nf ; Nf∗ ) ⊕ (Nf∗ ; Nf ) representation together with the A1 in the VM limit. We did not include the loop e/ects of the nucleon or constituent quarks which would become massless near the chiral restoration point. Inclusion of these would a/ect the result in this report. Such e/ects were studied by the ladder SD equation where the meson loop e/ects were ignored, 77

In the linear sigma model having a scalar meson (Higgs boson) in addition to the NG boson , the – scattering amplitude is expressed as A(s) = 6 + 2(6F )2 =(s − MS2 ), where MS2 = 26F2 with 6 being the four-point coupling. In the broken phase (F = 0) we can easily see that A(s = 0) = 0 consistently with the low-energy theorem, which holds even we approach the chiral restoration point (F → 0). In the symmetric phase (F ≡ 0), on the other hand, we have A(s = 0) = 6 = 0 which holds even at the low-energy limit s → 0: A(s = 0) = 0. In any case the amplitude is non-singular, although the low-energy limit amplitude is discontinuous across the phase transition point.

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instead. Since both approaches yield qualitatively the same result, there might exist some kind of duality between them. In this report we applied the VM to the chiral restoration in the large Nf QCD. It may be checked by the lattice simulation: As we obtained from a simple expectation around Eq. (6.118) and explicitly formulated in Section 6.3.2, the VM generally implies m2 →0 ; F2 (0)

(9.16)

which is a salient feature of the VM [106]. This will be a clear indication of the VM and may be testable in the lattice simulations. The results of Refs. [99,93] shown in Section 8 imply that the position of the peak of the dilepton spectrum will move to the lower energy region in accord with the picture shown in Refs. [42–45]. In the analysis we did not study the temperature dependence of the width. However, when the scaling properties of the parameters in hot QCD are equal to those in large >avor QCD, Eqs. (6.131), (6.135) and (6.136) would further imply smaller width and larger peak value near the critical point [see Eqs. (6.137) and (6.138)]: [106] 2 5=m ∼ g ∼ f(j) → 0 ;

(9.17)

2 m4 ) ∼ 1=f2 (j) → ∞ : 5ee 5 =52 ∼ g2 =(g

(9.18)

If it is really the case, these would be clear signals of VM tested in the future experiments. The VM reviewed in Section 6 may be applied to the models for the composite W and Z. Our analysis shows that the mass of the composite vector boson approaches to zero faster than the order parameter, which is 0xed to the electroweak symmetry breaking scale, near the critical point: m2 F2 (0) (250 GeV)2

(9.19)

in accord with the mass of W and Z bosons being smaller than 250 GeV. Moreover, near the VM point the composite theory becomes a weakly coupled gauge theory of the light gauge and NG bosons, while the underlying gauge theory is still in the strongly coupled phase with con0nement and chiral symmetry breaking. Such a situation has been hardly realized in the conventional strongly coupled dynamics for the composite gauge boson. The VM may also be applied to the technicolor with light techni-. In the present analysis we worked in the chiral limit with neglecting the e/ects from the current quark masses which explicitly break the chiral symmetry. For comparing the predictions for the system of the mesons other than the and such as K ∗ and K with experiment, we need to include the e/ects from the explicit breaking terms. Such analysis is also important for lattice analysis. In several analyses (see, e.g., Ref. [5]) where the chiral limit is usually taken by just the linear extrapolation. However, the chiral perturbation with systematically including the vector meson will generate the chiral logarithms in the chiral corrections to the vector meson masses. The chiral logarithms in the chiral perturbation theory in the light pseudoscalar meson system plays an important role, so that the inclusion of them in the chiral corrections to the vector meson masses is important to extrapolate the lattice results to the chiral limit. In conclusion we have developed an e/ective 0eld theory of QCD and QCD-like theories based on the HLS model. In contrast to other vector meson models which are all equivalent to the HLS

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model at tree level, we have provided a well-organized quantum 0eld theory and thus established a theory as a precise science which goes beyond a mere mnemonic of hadron phenomenology. In particular, we have presented a novel dynamical possibility for the chiral phase transition which is materialized through the quantum e/ects of the HLS model as the e/ective 0eld theory in such a way that the bare parameters of the HLS model are determined through matching with the underlying QCD-like theories. We do hope that it will shed some deeper insights into the strong coupling gauge theories and the concept of the composite gauge boson as well as the various possible phases of the hadronic matter. Acknowledgements We would like to thank Gerry Brown and Mannque Rho for useful discussions and continuous encouragements. We appreciate discussions with Tom Appelquist, Howard Georgi, Kazushi Kanaya, Yoshio Kikukawa, Youngman Kim, Taichiro Kugo, Ken Lane, Volodya Miransky, Chihiro Sasaki, Masaharu Tanabashi, Scott Thomas and Arkady Vainshtein. MH would like to thank Mannque Rho for his hospitality during the stay at KIAS, Gerry Brown for his hospitality during the stay at SUNY at Stony Brook and Dong-Pil Min for his hospitality during the stay at Seoul National University where part of this work was done. Part of this work was done when KY was staying at Aspen Center for Physics in 2002 summer. The work is supported in part by the JSPS Grant-in-Aid for Scienti0c Research (B) (2) 14340072 and 11695030. The work of MH was supported in part by USDOE Grant #DE-FG02-88ER40388 and the Brain Pool program (#012-1-44) provided by the Korean Federation of Science and Technology Societies. Appendix A. Convenient formulae A.1. Formulae for Feynman integrals Let us consider the following Feynman integrals:

1 dn k A0 (M 2 ) ≡ ; n 2 i(2) M − k 2

dn k 1 2 B0 (p ; M1 ; M2 ) ≡ ; 2 n 2 i(2) [M1 − k ] [M22 − (k − p)2 ]

k$ dn k $ ; B (p; M1 ; M2 ) ≡ i(2)n [M12 − k 2 ] [M22 − (k − p)2 ]

(2k − p)$ (2k − p) dn k $ : B (p; M1 ; M2 ) ≡ i(2)n [M12 − k 2 ] [M22 − (k − p)2 ]

(A.1) (A.2) (A.3) (A.4)

A0 (M 2 ) and B$ (p; M1 ; M2 ) are quadratically divergent. Since a naive momentum cuto/ violates the chiral symmetry, we need a careful treatment of the quadratic divergences. As discussed in Section 4, we adopt the dimensional regularization and identify the quadratic divergences with the presence of

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poles of ultraviolet origin at n = 2 [183]. This can be done by the following replacement in the Feynman integrals [see Eq. (4.85)]:

k $ k dn k 1 2 dn k 2 → ; → − g$ : (A.5) i(2)n −k 2 (4)2 i(2)n [ − k 2 ]2 2(4)2 As is usual, the logarithmic divergence is identi0ed with the pole at n = 4 by [see Eq. (4.99)] 2 1 − E + ln(4) + 1 → ln 2 ; +1≡ 4−n jS

(A.6)

where E is the Euler constant. Now, A0 (M 2 ) is evaluated as

2 M2 1 2 + 1 − ln M : − A0 (M ) = (4)2 (4)2 jS 2

(A.7)

B0 (p2 ; M1 ; M2 ) and B$ (p; M1 ; M2 ) are evaluated as

1 1 2 2 − F0 (p ; M1 ; M2 ) ; B0 (p ; M1 ; M2 ) = (4)2 jS

where

B$ (p; M1 ; M2 ) = p$ B1 (p2 ; M1 ; M2 ) ;

(A.8)

11 1 2 − F1 (p ; M1 ; M2 ) : B1 (p ; M1 ; M2 ) ≡ (4)2 2 jS

(A.9)

2

B$ (p; M1 ; M2 ) is evaluated as B$ (p; M1 ; M2 ) = −g$ [A0 (M1 ) + A0 (M2 ) − BA (p2 ; M1 ; M2 )] −(g$ p2 − p$ p )[B0 (p2 ; M1 ; M2 ) − 4B3 (p2 ; M1 ; M2 )] ; where

(A.10)

11 1 2 − F3 (p ; M1 ; M2 ) ; B3 (p ; M1 ; M2 ) ≡ (4)2 6 jS 2

BA (p2 ; M1 ; M2 ) ≡

1 (M12 − M22 )FA (p2 ; M1 ; M2 ) : (4)2

(A.11)

The de0nitions of F0 , FA , F1 and F3 and formulas are given in Appendix A.2. Here we summarize the divergent parts of the Feynman integrals which are used in Section 4.6 to obtain the divergent corrections to the parameters F , F and g: A0 (M 2 )|div =

2 M2 − ln 2 ; (4)2 (4)2

B0 (p2 ; M1 ; M2 )|div =

1 ln 2 ; (4)2

(A.12) (A.13)

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p$ ln 2 ; 2(4)2 1 B$ (p; M1 ; M2 )|div = −g$ [22 − (M12 + M22 ) ln 2 ] (4)2 B$ (p; M1 ; M2 )|div =

− (g$ p2 − p$ p )

1 ln 2 : 3(4)2

(A.14)

(A.15)

A.2. Formulae for parameter integrals Several parameter integrals are given as follows:

1 d x ln[(1 − x)M12 + xM22 − x(1 − x)s] ; F0 (s; M1 ; M2 ) = 0

FA (s; M1 ; M2 ) =

1

0

F1 (s; M1 ; M2 ) =

1

0

F2 (s; M1 ; M2 ) =

1

0

F3 (s; M1 ; M2 ) =

1

0

F4 (s; M1 ; M2 ) =

1

0

F5 (s; M1 ; M2 ) =

1

0

d x (1 − 2x) ln [(1 − x)M12 + xM22 − x(1 − x)s] ; d x x ln[(1 − x)M12 + xM22 − x(1 − x)s] ; d x x2 ln[(1 − x)M12 + xM22 − x(1 − x)s] ; d x x(1 − x) ln[(1 − x)M12 + xM22 − x(1 − x)s] ; d x ((1 − x)M12 + xM22 ) ln[(1 − x)M12 + xM22 − x(1 − x)s] ; d x ((1 − 2x)(M12 − M22 ) + (1 − 2x)2 s)

×ln[(1 − x)M12 + xM22 − x(1 − x)s] ; F6 (s; M1 ; M2 ) = F4 (s; M1 ; M2 ) − sF3 (s; M1 ; M2 ) : These are given by 2 2 S M1 ; M2 ) + M1 − M2 ln M1 − 2 + ln(M1 M2 ) ; F0 (s; M1 ; M2 ) = L(s; s M2

FA (s; M1 ; M2 ) = − F1 (s; M1 ; M2 ) =

M12 − M22 [F0 (s; M1 ; M2 ) − F0 (0; M1 ; M2 )] ; s

1 [F0 (s; M1 ; M2 ) − FA (s; M1 ; M2 )] ; 2

F2 (s; M1 ; M2 ) = F1 (s; M1 ; M2 ) − F3 (s; M1 ; M2 ) ;

(A.16)

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199

1 F3 (s; M1 ; M2 ) = F0 (s; M1 ; M2 ) 4 2(M12 + M22 ) 1 1− [F0 (s; M1 ; M2 ) − F0 (0; M1 ; M2 )] − 12 s

F4 (s; M1 ; M2 ) =

−

(M12 − M22 )2 [F0 (s; M1 ; M2 ) − F0 (0; M1 ; M2 ) − sF0 (0; M1 ; M2 )] 3s2

−

1 1 F0 (0; M1 ; M2 ) + ; 12 18

M12 + M22 M 2 − M22 F0 (s; M1 ; M2 ) + 1 FA (s; M1 ; M2 ) ; 2 2

F5 (s; M1 ; M2 ) = (M12 − M22 )FA (s; M1 ; M2 ) + sF0 (s; M1 ; M2 ) − 4sF3 (s; M1 ; M2 ) ; F6 (s; M1 ; M2 ) = F4 (s; M1 ; M2 ) − sF3 (s; M1 ; M2 ) ;

(A.17)

 2  −  (M1 + M2 )2 − s (M1 − M2 )2 − s   s      (M1 + M2 )2 − s + (M1 − M2 )2 − s   √ ×ln ;    2 M1 M2      (for s ¡ (M1 − M2 )2 ) ;              s − (M1 − M2 )2 2   ; (M1 + M2 )2 − s s − (M1 − M2 )2 × tan−1 s (M1 + M2 )2 − s S M1 ; M2 ) ≡ L(s;     (for (M1 − M2 )2 ¡ s ¡ (M1 + M2 )2 ) ;             2 s − (M + M )2 s − (M − M )2  1 2 1 2   s      2+ 2  s − (M + M ) s − (M − M ) 1 2 1 2   × ln √ − i ;   2 M M  1 2     (for (M1 + M2 )2 ¡ s) :

(A.18)

where

and F0 (0; M1 ; M2 ) =

M12 + M22 M1 ln − 1 + ln(M1 M2 ) ; M12 − M22 M2

(A.19)

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F0 (0; M1 ; M2 ) = −

M12 + M22 M12 M22 M12 + ln : 2(M12 − M22 )2 (M12 − M22 )3 M22

(A.20)

The following formulae are convenient: F0 (0; M; M ) = ln M 2 ; FA (0; M1 ; M2 ) = F3 (0; M; 0) =

(A.21)

M12 + M22 M12 M12 M22 ln ; − 2(M12 − M22 ) (M12 − M22 )2 M22

1 5 ln M 2 − : 6 36

(A.22) (A.23)

A.3. Formulae for generators Let me summarize useful formulae for the sum in terms the generators of SU(Nf ). In the following the generators are normalized as tr[Ta Tb ] = 12 ab :

(A.24)

The basic formulae for SU(Nf ) generators are given by Nf2 −1

tr[Ta ATa B] = −

a=1

1 1 tr[AB] + tr[A] tr[B] ; 2Nf 2

Nf2 −1

tr[Ta A] tr[Ta B] =

a=1

1 1 tr[A] tr[B] ; tr[AB] − 2 2Nf

(A.25)

(A.26)

where A and B are arbitrary Nf × Nf matrices. Below we list several convenient formulae for generators: Nf2 −1

a=1

Nf2 − 1 tr[ATa Ta ] = tr[A] ; 2Nf

Nf2 −1

tr[C[A; Ta ]] tr[D[B; Ta ]] =

a=1

1 tr[[C; A][D; B]] ; 2

(A.27)

(A.28)

Nf2 −1

tr[Ta [A; Tb ]] tr[Ta [B; Tb ]]

a;b=1

=−

Nf Nf 1 ˜ ; tr[AB] + tr[A] tr[B] = − tr[A˜ B] 2 2 2

(A.29)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233 Nf2 −1

tr[{A; B}{Ta ; Ta }] =

a=1 Nf2 −1

a=1

Nf2 −1

2(Nf2 − 1) tr[AB] ; Nf

201

(A.30)

Nf2 − 2 tr[{A; Ta }{B; Ta }] = tr[AB] + tr[A] tr[B] ; Nf

(A.31)

˜ ; tr[[A; Ta ][B; Ta ]] = −Nf tr[AB] + tr[A] tr[B] = −Nf tr[A˜ B]

(A.32)

a=1 Nf2 −1

tr [A{Ta ; Tb }] tr[B{Ta ; Tb }] =

a;b=1

Nf2 − 4 Nf2 + 2 tr[AB] + tr[A] tr[B] ; 2Nf 2Nf2

(A.33)

Nf2 −1

tr[C{Ta ; Tb }] tr[[A; Ta ] [B; Tb ]]

a;b=1

Nf 1 tr[{A; B}C] + (tr[A] tr[BC] + tr[B] tr[AC] − tr[C] tr[AB]) 4 2 Nf 1 ˜ B}C] ˜ ˜ ; tr[{A; − tr[C] tr[A˜ B] =− 4 2 where A˜ and B˜ are the traceless parts of A and B, respectively: =−

(A.34)

1 A˜ ≡ A − tr[A] ; Nf 1 B˜ ≡ B − tr[B] : Nf

(A.35)

A.4. Incomplete gamma function The incomplete gamma function is de0ned by

∞ d z −z j 5(j; A) ≡ e z : z A

(A.36)

For j = integer ¿ 1 these satisfy 5(1; A) = e−A ;

(A.37)

5( j ¿ 2; A) = e−A Aj−1 + (j − 1)5(j − 1; A) :

(A.38)

The incomplete gamma functions for j = 0 are approximately given by 1 : 5(0; A) ln A

(A.39)

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For j = integer ¡ 0 (j ¿ − 2) the incomplete gamma functions are given by 1 −A 1 1 ; 5(−1; A) = e − 5(0; A) − ln A A A

1 1 1 −A 1 1 1 + ln : 5(−2; A) = e − 5(−1; A) − 2 A2 2 A2 A A

(A.40) (A.41)

A.5. Polarization tensors at non-zero temperature In this subsection we list the polarization tensor at non-zero temperature, and give several convenient formulae among them. These polarization tensors are used in the calculation at non-zero temperature given in Section 8. At non-zero temperature, the polarization tensor is no longer restricted to be Lorentz covariant, but only O(3) covariant. Then the polarization tensors can be expressed by four independent symmetric O(3) tensors. Here we list the polarization tensors at non-zero temperature: [180,69] ˜j p ˜ ip PT$ = g$i ij − gj |˜ p|2    PT 00 = PT 0i = PTi0 = 0 ; = ˜j p ˜ ip  ;  PTij = ij − |˜ p|2 p $ p − PT$ PL$ ≡ − g$ − p2 p $ p0 p 2 p 0 p g0 − ; = g$0 − p2 |˜ p|2 p2 p$ p0 1 p 0 p p + p$ g0 − ; g$0 − PC$ ≡ √ p2 p2 2|˜ p| PD$ ≡

p $ p ; p2

where p$ = (p0 ; p ˜ ) is four-momentum. The following formulas are convenient: 78 PL$# PL# = −PL$ ; PT$# PT# = −PT$ ; PC$# PC# = 78

1 (P − PD$ ) ; 2 L$

There is an error in the third formula in Ref. [102].

(A.42)

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203

PD$# PD# = PD$ ; PL$# PT# = PC$# PT# = PD$# PT# = PD$# PL# = 0 ; p$ p0 p # # : g0 − PC$# PL = −PD$# PC = − √ p2 2|˜ p|

(A.43)

A.6. Functions used at non-zero temperature Here we list the functions used at non-zero temperature in Section 8. Functions used in the expressions of f in Eq. (8.4) are de0ned as follows:

∞ k n− 1 = I˜n T n ; dk k=T In (T ) ≡ e −1 0

∞ y n− 1 ˜I n = = (n − 1)!I(n) ; dy y e −1 0 2 I˜2 = ; 6

4 86 I˜4 = ; I˜6 = ; 15 63

∞ kn 1 Jmn (M ; T ) ≡ dk !=T ; e − 1 !m 0 ! ≡ k 2 + M2 :

n; m : integer ;

We also de0ne the functions in the -meson propagator as follows:

∞ 4k n 1 dk P !=T ; F3n (p0 ; M ; T ) ≡ e − 1 !(4!2 − p02 ) 0

∞ 4k 2 k n− 1 dk P k=T Gn (p0 ; T ) ≡ e − 1 4k 2 − p02 0

∞ p02 k n− 1 = In (T ) + dk P k=T ; e − 1 4k 2 − p02

∞0 kn 1 1 dk P !=T ; H1n (p0 ; M ; T ) ≡ 2 2 e − 1 ! (M − p0 )2 − 4k 2 p02 0

∞ 1 k n− 1 Kn (p0 ; M ; T ) ≡ dk P k=T ; 2 2 e − 1 (M − p0 )2 − 4k 2 p02 0

(A.44)

(A.45)

where P denotes the principal part. Appendix B. Feynman rules in the background 2eld gauge In this appendix we show the Feynman rules for the propagators of the quantum 0elds and the vertices including two quantum 0elds in the background 0eld gauge. The relevant Lagrangian is

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given in Eq. (4.84) in Section 4.4. In the following 0gures fabc is the structure constant of the SU(Nf ) group. Vertices with a dot (•) imply that the derivatives a/ect to the quantum 0elds, while those with a circle (◦) imply that no derivatives are included. B.1. Propagators Feynman rules for the propagators are shown in Fig. 23.

Fig. 23. Feynman rules for the propagators.

B.2. Three-point vertices Feynman rules for three-point vertices are shown in Figs. 24–26.

S $. Fig. 24. Feynman rules for the vertices which include one A

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

S$ . Fig. 25. Feynman rules for the vertices which include one V

Fig. 26. Feynman rules for the vertices which include one VS$ .

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B.3. Four-point vertices Feynman rules for four-point vertices are given in Figs. 27–32.

S $A S . Here summation over e is taken. Fig. 27. Feynman rules for the vertices which include A

S$ V S . Here summation over e is taken. Fig. 28. Feynman rules for the vertices which include V

S$ VS . Here summation over e is taken. Fig. 29. Feynman rules for the vertices which include V

Fig. 30. Feynman rules for the vertices which include VS$ VS . Here summation over e is taken.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

207

S . Here summation over e is taken. S $V Fig. 31. Feynman rules for the vertices which include A

S $ VS . Here summation over e is taken. Fig. 32. Feynman rules for the vertices which include A

Appendix C. Feynman rules in the Landau gauge In this appendix, for convenience, we show the Feynman rules for the propagators and the vertices in the Landau gauge with ordinary quantization procedure. The relevant Lagrangian is given in Eq. (4.20) in Section 4.2. The gauge 0xing is done by introducing an RC -gauge-like gauge-0xing and the corresponding Faddeev–Popov ghost Lagrangian is added [103]: 1 i LGF+FP = − tr[(9$ $ )] + agF2 tr[9$ $ (CL − C†L + CR − CR† )] # 2 1 1 † † 2 † † 2 2 2 4 (tr[CL − CL + CR − CR ]) + #a g F tr[(CL − CL + CR − CR ) ] − 16 Nf 1 † † $ 2 2 S + i tr C 29 D$ C + #ag F (CCL + CL C + CCR + CR C) ; 2

(C.1)

where # denotes a gauge parameter and C denotes a ghost 0eld. Here we choose the Landau gauge, # = 0. In this gauge the would-be NG boson is still massless, no other vector–scalar interactions are created and the ghost 0eld couples only to the HLS gauge 0eld $ . As in the Feynman rules for the background 0eld gauge in Appendix B, in the following 0gures fabc is the structure constant of the SU(Nf ) group. Vertices with a dot (•) imply that the derivatives are included, while those with a circle (◦) imply that no derivatives are included. For calculating the two-point functions at one-loop level, it is enough to have Feynman rules up until four-point vertices. In this appendix we do not list the vertices with more than four legs. It should be noticed that the Feynman rules in the Landau gauge of the covariant gauge 0xing used in Section 7 [see Eq. (7.5)] agree with those in the Landau gauge of the RC -gauge-like gauge-0xing listed below except for the ––– and ––– vertices. The term proportional to (p2 +p3 +p4 )$ for the ––– vertex as well as that for the ––– vertex [see Figs. 32(b) and (c)] does not exist in the Landau gauge of the covariant gauge 0xing.

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C.1. Propagators Feynman rules for the propagators in Landau gauge are given in Fig. 33.

Fig. 33. Feynman rules for the propagators in the Landau gauge.

C.2. Two-point vertices (mixing terms) Feynman rules in the Landau gauge for two-point vertices are given in Fig. 34.

Fig. 34. Feynman rules in the Landau gauge for two-point vertices (mixing terms).

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

C.3. Three-point vertices Feynman rules in the Landau gauge for the three-point vertices are given in Figs. 35–37.

Fig. 35. Feynman rules in the Landau gauge for three-point vertices which include one A$ .

Fig. 36. Feynman rules in the Landau gauge for the vertices which include one V$ .

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Fig. 37. Feynman rules in the Landau gauge for three-point vertices which include no V$ and A$ . Note that there are no –– and –– vertices.

C.4. Four-point vertices Feynman rules in the Landau gauge for four-point vertices are presented in Figs. 38–43.

Fig. 38. Feynman rules in the Landau gauge for four-point vertices which include one A$ . Here summations over e are taken. There is no A––– vertex.

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

211

Fig. 38. (continued).

Fig. 39. Feynman rules in the Landau gauge for four-point vertices which include one V$ . Here summations over e are taken. There are no V––– and V––– vertices.

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Fig. 40. Feynman rule for the four-point vertex which includes A$ A . Here summation over e is taken. Note that there are no A–A–– and A–A–– vertices.

S . Here summation over e is taken. Note that there S$ V Fig. 41. Feynman rule for the four-point vertex which includes V are no V–V–– and V–V–– vertices.

Fig. 42. Feynman rules in the Landau gauge for four-point vertices which include one but no V$ and A$ . Here summations over e are taken. There are no ––– and ––– vertices. Note that the second term proportional to (p2 + p3 + p4 )$ in (b) as well as in (c) does not exist in the Landau gauge of the covariant gauge 0xing.

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213

Fig. 43. Feynman rules in the Landau gauge for four-point vertices which include no , V$ and A$ . Here summations over e are taken.

Appendix D. Renormalization in the heat kernel expansion In this appendix we use the heat kernel expansion and the proper time regularization to determine the divergent contributions at one loop. D.1. Ghost contributions Let us 0rst consider the ghost contribution. The contribution is given by ˜ (CC) ; 5FP = −i Ln det ∇

(D.1)

˜ (CC) : ˜ (CC) ≡ (D˜ $ · D˜ $ )(CC) + M ∇ ab ab ab

(D.2)

where

By using the proper time integral with a ultraviolet cuto/ this 5FP is evaluated as

∞ dt S Tr 5FP = i (F(x; x; t))aa ; 1=2 t a

(D.3)

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S y; t))ab is obtained by solving the heat equation: where (F(x; 9 S ˜ (CC) S (F(x; y; t))ab + ∇ ac (F(x; y; t))cb = 0 : 9t c To solve the above heat equation let us start from the heat equation for the free 0eld: 9 S F 0 (x; y; t; M2 ) + ( + M2 )FS 0 (x; y; t; M2 ) = 0 : 9t The solution is given by

i (x − y)2 2 2 S − tM ; exp F 0 (x; y; t; M ) = (4t)n=2 4t

(D.4)

(D.5)

(D.6)

where n is the space–time dimension. The solution of Eq. (D.4) is expressed as S y; t))ab = FS 0 (x; y; t; M2 )(HS (x; y : t))ab ; (F(x; where (HS (x; y : t))ab satis0es 9 S (x − y)$ ˜ (CC) S (H (x; y; t))ab + (D$ )ac (H (x; y; t))cb 9t t c S + (D˜ $ · D˜ $ )(CC) ac (H (x; y; t))cb = 0 :

(D.7)

(D.8)

c

Eq. (D.8) can be solved by expanding (HS (x; y : t))ab in terms of t: (HS (x; y : t))ab = t j (HS j (x; y))ab :

(D.9)

j=0

Substituting this into Eq. (D.3) and taking n = 4 we obtain 1 S 2 2− j 5FP = − (M v ) 5(j − 2; A) Tr Hj (x; x) ; (4)2 j=0

(D.10)

where M2 ; 2 and 5(j; A) is the incomplete gamma function de0ned by

∞ d z −z j e z : 5(j; A) ≡ z A A=

(D.11)

(D.12)

Several convenient formulae for the incomplete gamma function are summarized in Appendix A.4. Eq. (D.8) is solved by substituting the expansion in Eq. (D.9). The results are given by (HS 0 (x; y))ab = ab

(normalization) ;

(D.13)

(HS 1 (x; y))ab = 0 ;

(D.14)

1 S (5$ · 5S $ )ab ; (HS 2 (x; y))ab = 12

(D.15)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

215

where ]ab : (5S $ )ab ≡ [D˜ $; (CC) ; D˜ (CC) $

(D.16)

By using the de0nitions in Eqs. (4.79) and (4.80) this (5S $ )ab is expressed as (5S $ )ab = 2i tr[VS$ [Ta ; Tb ]] : Now 5FP is evaluated as

Nf 1 4 4 $ ˜ ˜ 5(0; A) tr[V $ V ] + · · · ; d x −M 5(−2; A) + 5FP = (4)2 6

(D.17)

(D.18)

where dots stands for the non-divergent contributions. Using the formulae for the incomplete gamma function given in Appendix A.4, the divergent contribution to 5FP is evaluated as

Nf 1 2 4 S$ VS $ ] ; d x ln × (D.19) 5FP = tr [ V (4)2 M2 6 where we dropped the constant term. D.2. , V and contributions Let us calculate the one-loop contributions from , V and . These are given by 5PV =

i ˜ ; Ln Det ∇ 2

(D.20)

˜ is de0ned by where ∇ ˜ AB + O˜ AB : ˜ AB ≡ (D˜ $ · D˜ $ )AB + M ∇ Similarly to the ghost contributions this 5PV is evaluated as

∞ dt A i FA (x; x; t) ; 5PV = − Tr 2 1=2 t A

(D.21)

(D.22)

where (F(x; y; t))BA is obtained by solving the heat equation: 9 B ˜ CA FCB (x; y; t) = 0 : FA (x; y; t) + ∇ 9t C

(D.23)

This looks similar to the ghost case. However, it is much more diVcult to solve it since a di/erence appears in the heat equation in the free 0elds: 9 B ˜ CA )F0 BC (x; y; t) = 0 : F0 A (x; y; t) + (9CA + M 9t C

(D.24)

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The solution is given by F0 BA (x; y; t) = F˜ 0 (x; y; t)PAB ; where

  PAB =  

(D.25)



1

  ; 

2

e−tM

e and

(D.26)

−tM2

i (x − y)2 : exp (4t)n=2 4t

F˜ 0 (x; y; t) ≡

(D.27)

Hereafter we suppress the suVxes of the matrices. It is useful to express the full solution as S y; t) = F˜ 0 (x; y; t)[P · H (x; y; t)] ; F(x;

(D.28)

where H (x; y; t) satis0es (x − y)$ −1 ˜ 9 H (x; y; t) + P · D$ · P · H (x; y; t) 9t t ˜ · P · H (x; y; t) = 0 : +P −1 · (D˜ $ · D˜ $ + O)

(D.29)

Similarly to the ghost contributions Eq. (D.29) can be solved by expanding H (x; y; t) in terms of t: H (x; y; t) = t j Hj (x; y) : (D.30) j=0

First three, i.e., H0 , H1 and H2 are given by H0 (x; y) = 1;

(normalization) ;

(D.31)

H1 (x; y) = −P −1 · O˜ · P ; 1 1 ˜ ˜ 1 ˜$ ˜ ˜ −1 $ H2 (x; y) = P 5$ · 5 + O · O + [D ; [D$ ; O]] · P ; 12 2 6

(D.32) (D.33)

where 5$ ≡ [D˜ $ ; D˜ ] :

(D.34)

Then 5PV is formally evaluated by

∞ 1 dt t j−3 Tr (P · Hj (x; x)) : 5PV = 2(4)2 1=2 j=0

(D.35)

Since this expression includes an infrared divergence coming from the pion loops, we regularize this by introducing a small mass to pions. This is done by performing the following replacement:   −t$2 e   2  : (D.36) P → P˜ ≡  e−tM   2

e−tM

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

Let us evaluate 5PV step by step. The contribution for j = 0 is just a constant:

1 (0) d 4 x[5(−2; A) 5PV = ˜ + 5M4 5(−2; A)] ; 2(4)2

217

(D.37)

where M2 $2 ; A ˜ ≡ : 2 2 The contribution for j = 1 is given by

1 (1) () () 5PV = d 4 x[ − $2 5(−1; A)O ˜ aa − M2 5(−1; A)Oaa ] : 2(4)2 a A≡

(D.38)

(D.39)

Using the formulas given in Appendix A.3, we obtain Nf2 −1

() Oaa =

a=1

2 4 − 3a S $A S$ − VS$ )(V S $ − VS $ )] S $ ] + a Nf tr[(V Nf tr[A 2 2

+ Nf2 −1

() Oaa =

a=1

Nf2 − 1 F,2 tr[,S + ,S† ] ; Nf F2

a S $ ] + 1 Nf tr[(V S $A S$ − VS$ )(V S $ − VS $ )] : Nf tr[A 2 2

(1) is evaluated as By using the formulas for 5(j; A) in Appendix A.4, 5PV

1 2−a a 2 (1) 4 2 2 S $A S $] d tr[A = x − + M ln 5PV N N f f (4)2 2 4 M2 1 + a2 1 2 2 2 S$ − VS$ )(V S $ − VS $ )] + − Nf + Nf M ln 2 tr[(V 4 4 M Nf2 − 1 2 F,2 † − 2 tr[,S + ,S ] ; 2Nf F

(D.40) (D.41)

(D.42)

where we have taken $ = 0. (2) it is easy and enough to take $ = M , so that For identifying the logarithmic divergence in 5PV we can simply take Tr[H2 (x; x)] instead of Tr[P · H2 (x; x)]. Thus,

∞ 1 dt −tM2 (2) e Tr(H2 (x; x)) 5PV = 2 2(4) 1=2 t

1 1 1 4 $ ˜ ˜ tr[O · O] + tr[5$ · 5 ] ; = 5(0; A) d x (D.43) 2(4)2 2 12 where 5(0; A) ln

2 : M2

(D.44)

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* Let us calculate O˜ · O˜ parts by parts. a (O · O)() is given by aa () () (V ) ) () () (O · O)() [Oab Oba + Oab Oba + Oab # O(V# ba ] ; aa = a

a;b

(D.45)

#

where

() () Oab Oba

a;b

=

4 (4 − 3a)2 S $A S$ − VS$ )(V S $ − VS $ ))2 ] S $ )2 ] + a Nf tr[((V Nf tr[(A 8 8

+

a2 (4 − 3a) S $A S − VS )(V S − VS )] S $ (V Nf tr[A 4

+

2 (4 − 3a)2 S $A S $A S $A S $ ])2 + (4 − 3a) tr[A S ] tr[A S ] (tr[A 8 4

+

a4 S$ − VS$ )(V S $ − VS $ )])2 (tr[(V 8

+

a4 S − VS )] tr[(V S $ − VS $ )(V S − VS )] S$ − VS$ )(V tr[(V 4

+

a2 (4 − 3a) S $ ] tr[(V S − VS )(V S − VS )] S $A tr[A 4

+

a2 (4 − 3a) S − VS )] tr[A S $ (V S − VS )] S $ (V tr[A 2

+

F,2 F2 4 − 3a S $A S $A S $ (,S + ,S† )] + 4 − 3a , tr[A S $ ] tr[,S + ,S† ] Nf 2 tr[A 4 F 4 F2

+

F,2 a2 S$ − VS$ )(V S $ − VS $ )(,S + ,S† )] Nf 2 tr[(V 4 F

a2 F,2 S$ − VS$ )(V S $ − VS $ )] tr[,S + ,S† ] tr[(V 4 F2 2 2 Nf2 + 2 F,2 Nf2 − 4 F,2 † 2 tr[(,S + ,S ) ] + (tr[,S + ,S† ])2 ; + 8Nf F2 F2 8Nf2 +

() () Oab Oba

a;b

=

a(5a2 − 12a + 9) S $A S − VS )(V S $ − VS $ )] S (V Nf tr[A 8

(D.46)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

219

a2 (3 − 2a) S $ (V S − VS )A S (V S $ − VS $ )] Nf (tr[A 8 S $ − VS $ )A S (V S − VS )]) S $ (V + tr[A

+

+

a(9 − 6a + a2 ) S $ (V S $ − VS $ )] tr[A S (V S − VS )] tr[A 8

+

a(9 − 6a + a2 ) S $A S $ − VS $ )(V S − VS )] S ] tr[(V tr[A 8

+

a(9 − 6a + a2 ) S $ (V S − VS )] tr[(V S $ − VS $ )A S ] tr[A 8

F,2 3a(a − 1) S $; V S $ − VS $ ]] Nf 2 tr[(,S − ,S† )[A 16 F 2 F,2 1 a Nf (tr[(,S − ,S† )2 ] − (tr[,S − ,S† ])2 ) ; − 2 32 F Nf +

#

a;b

(V# ) ) S $] : S $A Oab O(V# ba = −2aM2 Nf tr[A

(D.47) (D.48)

* () () Oba we used the form in Eq. (4.70), It should be noticed that in the above expression for a; b Oab which was rewritten by using the equation of motion (4.51). * Next a (O · O)() is given by aa () () (V ) ) () () () # Oab Oba + Oab Oba + (D.49) (O · O)aa = Oab O(V# ba ; a

#

a;b

where

() () Oab Oba =

a;b

() () Oab Oba ;

a;b

() () Oab Oba

a;b

=

a2 S $ )2 ] + 1 Nf tr[((V S $A S$ − VS$ )(V S $ − VS $ ))2 ] Nf tr[(A 8 8 a S $ (V S $A S − VS )(V S − VS )] + Nf tr[A 4 +

2 a2 S $A S $A S $ ])2 + a tr[A S ] tr[A S ] S $A (tr[A 8 4

(D.50)

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M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

+

1 S$ − VS$ )(V S $ − VS $ )])2 (tr[(V 8

1 S − VS )] tr[(V S $ − VS $ )(V S − VS )] S$ − VS$ )(V + tr[(V 4 a S $A S − VS )(V S − VS )] S $ ] tr[(V + tr[A 4 a S $ (V S − VS )] tr[A S $ (V S − VS )] ; + tr[A 2 (V ) ) S$ − VS$ )(V S $ − VS $ )] : Oab # O(V# ba = −2M2 Nf tr[(V a;b

(D.51) (D.52)

#

* *

V# ) (O · O)(V is given by # aa   (V V ) V# ) O) O(V# ) + O) O(V# ) + (O · O)(V = O(V# VL )ab ObaL #  ; (V# ab ba (V# ab ba # aa

Next

a

a

where

#

#

a;b

#

a;b

#

a;b

#

a;b

) O(V O(V# ) = # ab ba ) O(V O(V# ) = # ab ba

(V V# )

O(V# VL )ab ObaL

#

a;b

#

a;b

(D.53)

L

(V# ) ) S $A S $] ; Oab O(V# ba = −2aM2 Nf tr[A

(D.54)

(V# ) ) S$ − VS$ )(V S $ − VS $ )] ; Oab O(V# ba = −2M2 Nf tr[(V

(D.55)

= 8Nf tr[VS$ VS $ ] :

(D.56)

#;L

Summing over the above

* a

(O · O)() aa ,

* a

(O · O)() and aa

*

*

a; b

#

) O(V O(V# ) , we obtain # ab ba

1 ˜ tr(O˜ · O) 2 1 () () 1 () () 1 (V V ) = Oab Oba + Oab Oba + O(V# VL )ab ObaL # 2 2 2 a;b

+

a;b

a;b

() () Oab Oba +

a;b

#

a;b

(V# ) ) Oab O(V# ba +

#;L

a;b

#

(V# ) ) Oab O(V# ba

S $ ] − 2M2 Nf tr[(V S $A S$ − VS$ )(V S $ − VS $ )] =4Nf tr[VS$ VS $ ] − 2aM2 Nf tr[A +

4 5a2 − 12a + 8 S $A S$ − VS$ )(V S $ − VS $ ))2 ] S $ )2 ] + a + 1 Nf tr[((V Nf tr[(A 8 16

+

a(1 + 4a − 3a2 ) S $A S − VS )(V S − VS )] S $ (V Nf tr[A 8

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

+

221

a(5a2 − 12a + 9) S $A S − VS )(V S $ − VS $ )] S (V Nf tr[A 8

a2 (3 − 2a) S $ (V S − VS )A S (V S $ − VS $ )] Nf (tr[A 8 S $ (V S $ − VS $ )A S (V S − VS )]) + tr[A

+

+

2 5a2 − 12a + 8 S $A S $A S $A S $ ])2 + 5a − 12a + 8 tr[A S ] tr[A S ] (tr[A 8 4

+

a4 + 1 S$ − VS$ )(V S $ − VS $ )])2 (tr[(V 16

+

a4 + 1 S$ − VS$ )(V S − VS )] tr[(V S $ − VS $ )(V S − VS )] tr[(V 8

+

a(1 + 4a − 3a2 ) S $A S $ ] tr[(V S − VS )(V S − VS )] tr[A 8

+

a(9 − 6a + a2 ) S $A S $ − VS $ )(V S − VS )] S ] tr[(V tr[A 8

+

a(9 − 6a + a2 ) S $ − VS $ )])2 S $ (V (tr[A 8

+

a(1 + 4a − 3a2 ) S − VS )] tr[A S $ (V S − VS )] S $ (V tr[A 4

+

a(9 − 6a + a2 ) S − VS )] tr[(V S $ − VS $ )A S ] S $ (V [tr A 8

+

F,2 F2 4 − 3a S $ (,S + ,S† )] + 4 − 3a , tr[A S $ ] tr[,S + ,S† ] S $A S $A Nf 2 tr[A 8 F 8 F2

F,2 a2 S$ − VS$ )(V S $ − VS $ )(,S + ,S† )] + Nf 2 tr[(V 8 F +

a2 F,2 S$ − VS$ )(V S $ − VS $ )] tr[,S + ,S† ] tr[(V 8 F2

F,2 3a(a − 1) S $; V S $ − VS $ ]] Nf 2 tr[(,S − ,S† )[A 16 F 2 2 Nf2 − 4 F,2 Nf2 + 2 F,2 † 2 + tr[(,S + ,S ) ] + (tr[,S + ,S† ])2 16Nf F2 F2 16Nf2 +

a − Nf 32

F,2 F2

2 (tr[(,S − ,S† )2 ] −

1 (tr[,S − ,S† ])2 ) ; Nf

(D.57)

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Next let us calculate 5$ de0ned in Eq. (D.34): S S 5$ () ab = i tr[(aV$ + (2 − a)V$ )[Ta ; Tb ]]

a(2 − a) S 4 − 3a S S S S S [V$ − V$ ; V − V ] + [A$ ; A ] [Ta ; Tb ] ; − tr 2 2

√ 1 S 1 S S () S S S S 5$ab = a tr i A$ − [V$ − V$ ; A ] − [A$ ; V − V ] [Ta ; Tb ] 2 2

(D.58)

√ a−1 S ; Tb ]] S$ − VS$ ; Ta ][A tr[[V − a 2 √ a−1 S − VS ; Ta ][A S $ ; Tb ]] ; tr[[V + a 2

√ S$ − VS$ ; A S $; V S $ − 1 [V S ] − 1 [A S − VS ] [Ta ; Tb ] = a tr i A 5$() ab 2 2

(D.59)

√ a−1 S − VS ; Tb ]] S $ ; Ta ][V − a tr[[A 2 √ a−1 S$ − VS$ ; Tb ]] ; S ; Ta ][V + a tr[[A 2

(D.60)

S$ )[Ta ; Tb ]] 5$() = tr[(i VS$ + i V ab −

1 S$ − VS$ ; V S − VS ] + (2 − a)[A S $; A S ])[Ta ; Tb ]] ; tr[([V 2

V )

5$ (V#L = 2i tr[VS$ [Ta ; Tb ]]g#L ; ab

where we used the relations in Eqs. (4.31) and (4.32). Using this we obtain a2 1 () $() a(2 − a) S $ ] Nf tr[VS$ V 5$ab 5 ba = − Nf tr[VS$ VS $ ] − 12 24 12 a;b

−

(2 − a)2 S $ ] − i a(4 − 3a) Nf tr[VS$ A S $A S ] S$ V Nf tr[V 24 12

−i

a2 (2 − a) S $ − VS $ )(V S − VS )] Nf tr[VS$ (V 12

−i

(4 − 3a)(2 − a) S $A S ] S$ A Nf tr[V 12

−i

a(2 − a)2 S$ (V S $ − VS $ )(V S − VS )] Nf tr[V 12

−

2 (4 − 3a)2 S $A S $A S $ )2 ] + (4 − 3a) Nf tr[A S A S $A S ] Nf tr[(A 48 48

(D.61) (D.62)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

−

a2 (2 − a)2 S$ − VS$ )(V S $ − VS $ ))2 ] Nf tr[((V 48

+

a2 (2 − a)2 S$ − VS$ )(V S − VS )(V S $ − VS $ )(V S − VS )] Nf tr[(V 48

+

a(2 − a)(4 − 3a) S $A S $ − VS $ )(V S − VS )] S (V Nf tr[A 24

−

a(2 − a)(4 − 3a) S $A S − VS )(V S $ − VS $ )] ; S (V Nf tr[A 24

223

(D.63)

1 1 1 () $() S$ V S $ ] − 1 Nf tr[V S $ ] Nf tr[VS$ V 5$ab 5 ba = − Nf tr[VS$ VS $ ] − 12 24 12 24 a;b

2−a S $A S ] − i 1 Nf tr[VS$ (V S $ − VS $ )(V S − VS )] Nf tr[VS$ A 12 12 2−a S $A S ] − i 1 Nf tr[V S$ A S$ (V S $ − VS $ )(V S − VS )] Nf tr[V −i 12 12

−i

−

2 (2 − a)2 S $A S $A S $ )2 ] + (2 − a) Nf tr[A S A S $A S ] Nf tr[(A 48 48

−

1 S$ − VS$ )(V S $ − VS $ ))2 ] Nf tr[((V 48

+

1 S$ − VS$ )(V S − VS )(V S $ − VS $ )(V S − VS )] Nf tr[(V 48

2−a S $A S $ − VS $ )(V S − VS )] S (V Nf tr[A 24 2−a S $A S − VS )(V S $ − VS $ )] ; S (V Nf tr[A − 24 +

2 1 VL ) 5$(V# 5$ V# ) = − Nf tr[VS$ VS $ ] ; (VL ab 3 12 ba a;b

#;L

1 () $() a S $ A S $ [A S $; V S − VS ]] S $ ] − i a(a + 1) Nf tr[A 5$ab 5 ba = − Nf tr [A 6 12 12 a;b

−

a(a2 + 1) S $A S − VS )(V S − VS )] S $ (V Nf tr[A 24

+

a(a2 + 1) S $A S $ − VS $ )(V S − VS )] S (V Nf tr[A 24

(D.64)

(D.65)

224

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

a2 S $ (V S $ − VS $ )A S (V S − VS )] Nf (tr[A 24 S $ (V S − VS )A S ]) S$ − VS$ )A +tr[(V −

+

a2 S $ (V S − VS )A S $ (V S − VS )] Nf tr[A 12

−

a(a − 1)2 S $ ]tr[(V S $A S − VS )(V S − VS )] Nf tr[A 24

+

a(a − 1)2 S $A S $ − VS $ )(V S − VS )] S ]tr[(V Nf tr[A 24

+

a(a − 1)2 S $ (V S $ − VS $ )])2 Nf (tr[A 24

−

a(a − 1)2 S $ (V S − VS )]tr[A S $ (V S − VS )] Nf tr[A 12

+

a(a − 1)2 S $ (V S − VS )]tr[(V S $ − VS $ )#S⊥ ] : Nf tr[A 24

Thus, we obtain 1 tr(5$ · 5$ ) 12 1 () $ V ) () $() () $ba () [5$ab 5ba() + 5$ 5 ba + 2 5$ 5 + 5$ (VL# 5$ V# ) ] = ab ab (VL ab 12 ba a;b

17 + a2 1 + 2a − a2 S $ ] Nf tr[VS$ VS $ ] − Nf tr[VS$ V 24 12

=− −

5 − 4a + a2 S$ V S $ A S $ ] − a Nf tr[A S $ ] Nf tr[V 12 24

−i

2 + 3a − 3a2 S $A S ] Nf tr[VS$ A 12

−i

1 + 2a2 − a3 S $ − VS $ )(V S − VS )] Nf tr[VS$ (V 12

−i

(2 − a)(5 − 3a) S $A S ] S$ A Nf tr[V 12

−i

1 + 4a − 4a2 + a3 S$ (V S $ − VS $ )(V S − VS )] Nf tr[V 12

(D.66)

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

−i

225

a(a + 1) S $ [A S $; V S − VS ]] Nf tr[A 12

−

2 10 − 14a + 5a2 S $ )2 ] + 10 − 14 + 5a Nf tr[A S A S $A S ] S $A S $A Nf tr[(A 24 24

−

1 + 4a2 − 4a3 + a4 S$ − VS$ )(V S $ − VS $ ))2 ] Nf tr[((V 48

+

1 + 4a2 − 4a3 + a4 S$ − VS$ )(V S − VS )(V S $ − VS $ )(V S − VS )] Nf tr[(V 48

−

a(a2 + 1) S $ (V S $A S − VS )(V S − VS )] Nf tr[A 24

+

1 + 4a − 5a2 + 2a3 S (V S $A S $ − VS $ )(V S − VS )] Nf tr[A 12

−

(2 − a)(1 + 4a − 3a2 ) S $A S − VS )(V S $ − VS $ )] S (V Nf tr[A 24

−

a2 S $ (V S $ − VS $ )A S (V S − VS )] + tr[(V S$ − VS$ )A S $ (V S − VS )A S ]) Nf (tr[A 24

+

a2 S $ (V S − VS )A S $ (V S − VS )] Nf tr[A 12

−

a(a − 1)2 S $ ]tr[(V S $A S − VS )(V S − VS )] Nf tr[A 24

+

a(a − 1)2 S ] tr[(V S $A S $ − VS $ )(V S − VS )] Nf tr[A 24

+

a(a − 1)2 S $ (V S $ − VS $ )])2 Nf (tr[A 24

−

a(a − 1)2 S $ (V S − VS )] tr[A S $ (V S − VS )] Nf tr[A 12

+

a(a − 1)2 S $ (V S − VS )]tr[(V S $ − VS $ )A S ] : Nf tr[A 24

Finally, 1 S + 1 tr(5$ · 5$ ) tr(OS · O) 2 12 S $A S$ − VS$ )(V S $ − VS $ )] S $ ] − 2M2 Nf tr[(V = − 2aM2 Nf tr[A

(D.67)

226

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

+

79 − a2 5 − 4a + a2 S$ V S $ ] Nf tr[VS$ VS $ ] − Nf tr[V 24 24

−

2 a S $ ] − 1 + 2a − a Nf tr[VS$ V S $ ] S $ A Nf tr[A 12 12

−i

2 + 3a − 3a2 S $A S ] Nf tr[VS$ A 12

1 + 2a2 − a3 S $ − VS $ )(V S − VS )] Nf tr[VS$ (V 12 (2 − a)(5 − 3a) S $A S ] S$ A Nf tr[V −i 12

−i

1 + 4a − 4a2 + a3 S$ (V S $ − VS $ )(V S − VS )] Nf tr[V 12 a(a + 1) S $ [A S $; V S − VS ]] Nf tr[A −i 12 −i

+

2 7 − 11a + 5a2 S $A S $A S $ )2 ] + 10 − 14 + 5a Nf tr[A S A S $A S ] Nf tr[(A 12 24

+

1 − 2a2 + 2a3 + a4 S$ − VS$ )(V S $ − VS $ ))2 ] Nf tr[((V 24

+

1 + 4a2 − 4a3 + a4 S$ − VS$ )(V S − VS )(V S $ − VS $ )(V S − VS )] Nf tr[(V 48

+

a(1 + 6a − 5a2 ) S $A S − VS )(V S − VS )] S $ (V Nf tr[A 12

+

1 + 4a − 5a2 + 2a3 S $A S $ − VS $ )(V S − VS )] S (V Nf tr[A 12

−

1 − 10a + 13a2 − 6a3 S $A S − VS )(V S $ − VS $ )] S (V Nf tr[A 24

a2 (4 − 3a) S $ (V S $ − VS $ )A S (V S − VS )] Nf (tr[A 12 S $ (V S − VS )A S ]) S$ − VS$ )A + tr[(V +

+

a2 S $ (V S − VS )A S $ (V S − VS )] Nf tr[A 12

+

2 8 − 12a + 5a2 S $A S $A S $ ])2 + 8 − 12a + 5a tr[A S ] tr[A S ] S $A (tr[A 8 4

+

1 + a4 S$ − VS$ )(V S $ − VS $ )])2 (tr[(V 16

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

+

1 + a4 S$ − VS$ )(V S − VS )] tr[(V S $ − VS $ )(V S − VS )] tr[(V 8

+

a(1 + 7a − 5a2 ) S $A S − VS )(V S − VS )] S $ ] tr[(V tr[A 12

+

a(7 − 5a + a2 ) S $A S $ − VS $ )(V S − VS )] S ] tr[(V tr[A 6

+

a(7 − 5a + a2 ) S $ (V S $ − VS $ )])2 (tr[A 6

+

a(1 + 7a − 5a2 ) S $ (V S − VS )] tr[A S $ (V S − VS )] tr[A 6

+

a(7 − 5a + a2 ) S $ (V S − VS )] tr[(V S $ − VS $ )A S ] tr[A 8

+

F,2 F2 4 − 3a S $A S $A S $ (,S + ,S† )] + 4 − 3a , tr[A S $ ] tr[,S + ,S† ] Nf 2 tr[A 8 F 8 F2

+

F,2 a2 S$ − VS$ )(V S − VS )(,S + ,S† )] Nf 2 tr[(V 8 F

+

a2 F,2 S$ − VS$ )(V S − VS )] tr[,S + ,S† ] tr[(V 8 F2

227

F,2 3a(a − 1) S $; V S $ − VS $ ]] Nf 2 tr[(,S − ,S† )[A 16 F 2 2 Nf2 + 2 F,2 Nf2 − 4 F,2 † 2 tr[(,S + ,S ) ] + (tr[,S + ,S† ])2 + 16Nf F2 F2 16Nf2

+

a − Nf 32

F,2 F2

2 (tr [(,S − ,S† )2 ] −

1 (tr[,S − ,S† ])2 ) : Nf

(D.68)

(1) From the above lengthy equation, 5FP in Eq. (D.19) and 5PV in Eq. (D.42) we can read 2 the divergent corrections to the parameters at O(p ). Together with the bare parameters they are given by Nf 2 − a 2 3a 2 2 M − ; (D.69) F2 + − ln (4)2 2 4 M2

F2

Nf + (4)2

1 + a2 2 3 2 2 − M ln 2 − 4 4 M

;

(D.70)

228

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

Table 20 CoeVcients of the divergent corrections to zi in Eq. (4.27), wi in Eq. (4.26) and yi in Eq. (4.25); the normalization is 0xed by requiring that zi + (5zi =(4(4)2 )) ln(2 =M2 ) is 0nite z1

−Nf

5 − 4a + a2 12

y1

Nf

7 − 11a + 5a2 6

z2

−Nf

a 6

y2

Nf

10 − 14a + 5a2 12

z3

−Nf

1 + 2a − a2 6

y3

Nf

1 − 2a2 + 2a3 + a4 12

z4

−Nf

2 + 3a − 3a2 6

y4

Nf

1 + 4a2 − 4a3 + a4 24

z5

−Nf

1 + 2a2 − a3 6

y5

Nf

a(1 + 6a − 5a2 ) 6

z6

−Nf

(2 − a)(5 − 3a) 6

y6

Nf

1 + 4a − 5a2 + 2a3 6

z7

−Nf

1 + 4a − 4a2 + a3 6

y7

−Nf

z8

Nf

y8

Nf

a2 (4 − 3a) 6

y9

Nf

a2 6

y10

8 − 12a + 5a2 4

y11

8 − 12a + 5a2 2

y12

1 + a4 8

y13

1 + a4 4

y14

a(1 + 7a − 5a2 ) 6

a(a + 1) 6

4 − 3a 4

1 − 10a + 13a2 − 6a3 12

w1

Nf

w2

4 − 3a 4

w3

Nf

w4

a2 4

w5

−Nf

w6

Nf2 − 4 8Nf

y15

a(7 − 5a + a2 ) 3

w7

Nf2 + 2 8Nf

y16

a(7 − 5a + a2 ) 3

w8

−Nf

a 16

y17

a(1 + 7a − 5a2 ) 3

w9

a 16

y18

a(7 − 5a + a2 ) 3

a2 4

3a(a − 1) 8

M. Harada, K. Yamawaki / Physics Reports 381 (2003) 1 – 233

−

Nf 87 − a2 2 1 ln + ; 2g2 (4)2 48 M2

F,2 Nf2 − 1 2 − F2 : 4 2Nf (4F )2 ,

229

(D.71) (D.72)

The logarithmically divergent corrections to the coeVcients of O(p4 ) terms are listed in Table 20. Here the normalization is 0xed by requiring that zi + (5zi =(4(4)2 )) ln(2 =M2 ) is 0nite. [The normalizations for 5yi and 5wi are de0ned in the same way.] References [1] [2] [3] [4] [5] [6] [7] [8] [9] [10] [11] [12] [13] [14] [15] [16] [17] [18] [19] [20] [21] [22] [23] [24] [25] [26] [27] [28] [29] [30] [31] [32] [33] [34] [35]

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Available online at www.sciencedirect.com

Physics Reports 381 (2003) 235 – 402 www.elsevier.com/locate/physrep

Strong dynamics and electroweak symmetry breaking Christopher T. Hilla;∗ , Elizabeth H. Simmonsb; c a

Fermi National Accelerator Laboratory, P.O. Box 500, Batavia, IL 60510, USA Department of Physics, Boston University, 590 Commonwealth Avenue, Boston, MA 02215, USA c Radcli,e Institute for Advanced Study and Department of Physics, Harvard University, Cambridge, MA 02138, USA b

Accepted 6 March 2003 editor: J.A. Bagger

Abstract

√ The breaking of electroweak symmetry, and origin of the associated “weak scale”, vweak = 1= 2 2GF = 175 GeV, may be due to a new strong interaction. Theoretical developments over the past decade have led to viable models and mechanisms that are consistent with current experimental data. Many of these schemes feature a privileged role for the top quark, and third generation, and are natural in the context of theories of extra space dimensions at the weak scale. We review various models and their phenomenological implications which will be subject to de
PACS: 12.15.−y; 11.30.Qc

Contents 1. Introduction . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.1. Lessons from QCD . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.2. The weak scale . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.3. Superconductors, chiral symmetries, and Nambu–Goldstone bosons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.4. The standard model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.5. Purpose and synopsis of the review . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2. Technicolor . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.1. Dynamics of technicolor . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.1.1. The TC ↔ QCD analogy . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . ∗

Corresponding author. E-mail address: [email protected] (C.T. Hill).

c 2003 Published by Elsevier B.V. 0370-1573/03/$ - see front matter doi:10.1016/S0370-1573(03)00140-6

237 237 238 240 245 248 250 250 251

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2.1.2. Estimating in TC by rescaling QCD; f , FT , vweak . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2. The minimal TC model of Susskind and Weinberg . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2.1. Structure . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2.2. Spectroscopy of the minimal model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2.3. Non-resonant production and longitudinal gauge boson scattering . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2.4. Techni-vector meson production and vector meson dominance . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.3. Farhi–Susskind model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.3.1. Structure . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.3.2. Spectroscopy . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.3.3. Production and detection at hadron colliders . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.3.4. Production and detection at e+ e− colliders . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3. Extended technicolor . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1. The general structure of ETC . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1.1. Master gauge group GETC . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1.2. Low energy relic interactions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1.3. The -terms: techniaxion masses . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1.4. The terms: quark and lepton masses . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1.5. The terms: Iavor-changing neutral currents . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2. Oblique radiative corrections . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3. Some explicit ETC models . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3.1. Techni-GIM . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3.2. Non-commuting ETC models . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3.3. Tumbling and triggering . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3.4. Grand uni
252 254 254 255 261 262 266 266 267 269 277 283 283 283 285 286 287 289 290 294 294 295 297 297 298 298 299 302 305 306 313 314 317 319 321 321 322 327 327 327 327 331 331 332 336 338 341 341 344 347 351

C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 4.4.1. The minimal model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.4.2. Dynamical issues . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.4.3. Including the b-quark . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5. Top Seesaw phenomenology . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5.1. Seesaw quarks . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5.2. Flavorons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.6. Extra dimensions at the TeV scale . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.6.1. Deconstruction . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.6.2. Little Higgs theories . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5. Outlook and conclusions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Acknowledgements . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix A. The standard model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix B. The Nambu–Jona-Lasinio model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . References . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

237 352 353 357 359 359 365 367 368 372 374 375 375 382 385

1. Introduction 1.1. Lessons from QCD The early days of accelerator-based particle physics were largely explorations of the strong interaction scale, associated roughly with the proton mass, of order 1 GeV. Key elements were the elaboration of the hadron spectroscopy; the measurements of cross-sections; the elucidation of spontaneously broken chiral symmetry, with the pion as a Nambu–Goldstone phenomenon [1–3]; the evolution of the Iavor symmetry SU (3) and the quark model [4–6]; the discovery of scaling behavior in electroproduction [7–9]; and the observation of quarks and gluons as partons. Eventually, this work culminated in quantum chromodynamics (QCD): a description of the strong scale based upon the elegant symmetry principle of local gauge invariance, and the discovery of the Yang–Mills gauge group SU (3) of color [10,11]. Today we recognize that the strong scale is a well-de
Proponents of concepts such as nuclear democracy, duality, and Veneziano models (which contain an underlying string theory spectrum) believed this perspective, at the time, to be fundamental.

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mechanism leading to chiral symmetry breaking [1–3]. Finally, the strong interactions are readily visible in nature for what might be termed “contingent” reasons: the nucleon is stable, and atomic nuclei are abundant. If a process like p → e+ + occurred with a large rate, so that protons were short-lived, then the strong interactions would be essentially decoupled from low energy physics, √ for s ¡ 2m . A new strong dynamics, if it exists, presumably does not have an analogous stable sector (else we would have seen it), and this dynamics must be largely decoupled below threshold. QCD provides direct guidance for this review because of the light it may shed on the scale of electroweak symmetry breaking (EWSB). The origin of the scale QCD , and the “hierarchy” between the strong and gravitational scales is, in principle, understood within QCD in a remarkable and compelling way. If a perturbative input for s is speci<ed at some high energy scale, e.g., the Planck scale, then the logarithmic renormalization group running of s naturally produces a strong-interaction scale QCD which lies far below MPlanck . The value of scale QCD does not derive directly from the Planck scale, e.g. through multiplication by some ratio of coupling constants. Rather QCD arises naturally and elegantly from quantum mechanics itself, through dimensional transmutation [17]. Hence, QCD produces the large hierarchy of scales without
At this writing there are internal inconsistencies in the precision electroweak data; the Z 0 -pole data of leptons alone predicts a Higgs boson mass that is ∼ 20–40 GeV and is directly ruled out, while the hadronic AbFB predicts a very heavy Higgs mass ∼ 300 GeV; these discrepancies are signi
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239

at the TeV scale, the Higgs boson could be a bound state and the upper bound on the composite Higgs mass would rise to ∼ 1 TeV [24]. Certain models we will describe in this review, such as the Top Quark Seesaw scheme, predict a composite Higgs boson with a mass mH ∼ 1 TeV and are otherwise in complete agreement with electroweak constraints (see Section 4). In considering the Higgs sector, the foremost question is that of motive: “why should nature provide a unique elementary particle simply for the purpose of breaking a symmetry?” Other issues involve naturalness, i.e., the degree of
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systematically Iawed. Thus, while many of the fundamental symmetries controlling the known forces in nature are understood, speculation as to what lies on energy scales well above vweak should be viewed as tentative. Let us therefore focus on physics at the weak scale. We will begin by providing an introductory tutorial survey of the elementary physical ingredients of dynamical symmetry breaking. In Section 1.3, we will consider a sequence of “
3

In a superconductor the mass gap is actually a small Majorana-mass, ∼ carries net charge ±2.

+ h:c: for an electron, an operator which

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241

(i) Superconductor ↔ a massive photon The de
(1.1)

In the second line we have completed the square of the scalar terms and we see that this Lagrangian is gauge invariant, i.e., if the transformation A% → A% + 9% &=ef is accompanied by ' → ' + & then L is invariant. The quantity f, is called the “decay constant” of '. It is the direct analogue of f for the pion of QCD (for a discussion of normalization conventions for f see Section 2.1). We see that the photon vector potential and the massless mode have combined to form a new <eld: B% = A% − 9% '=ef. Physically, B% corresponds to a “gauge-invariant massive photon” of mass m = ef. Thus, the Lagrangian can be written directly in terms of B% as L = − 14 FB%( FB%( + 12 m2 (B% )2 ;

(1.2)

where FB%( ≡ 9% B( − 9( B% = F%( . The ' <eld has now blended with A% to form the heavy photon <eld; we say that the <eld ' has been “eaten” by the gauge <eld to give it mass. More generally, in order for this mechanism to produce a superconductor, the Lagrangian for the mode ' must possess the symmetry ' → ' + )(x) where )(x) can be any function of space–time (' → ' + ) where ) is a constant is the corresponding global symmetry in the absence of the gauge <elds). This essentially requires a massless <eld ', with derivative couplings to conserved currents ∼ j% 9% '. The shift will then change the ' action by at most a total divergence, and we can eliminate surface terms by requiring that all <elds be well behaved at in
= 12 (1 − 5 ) ;

R

= 12 (1 + 5 )

:

(1.3)

The (1 ± 5 )=2 operators are projections, and the reduced <elds are equivalent to two independent two-component complex spinors, each, by itself, forming an irreducible representation of the Lorentz group. The Lagrangian of a massless Dirac spinor decomposes into two independent <elds’ kinetic terms as L = U i = U Li

L

+ U Ri

R

:

(1.4)

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This Lagrangian is invariant under two independent global symmetry transformations, which we call the “chiral symmetry” U (1)L × U (1)R : L

→ exp(−i,) L ;

R

→ exp(−i!)

R

:

(1.5)

The symmetry transformation corresponding to the conserved fermion number has (, = !) while an axial, or 5 , symmetry transformation has (, = −!). The corresponding Noether currents are j%L ≡

-L 1 = U % (1 − 5 ) ; -9% ,(x) 2

j%R ≡

-L 1 = U % (1 + 5 ) -9% !(x) 2

:

(1.6)

We can form the vector current, j% = j%R + j%L = U % and the axial vector current, j%5 = j%R − j%L = U % 5 U . If we add a mass term to our Lagrangian we couple together the two independent L- and R-handed <elds and thus break the chiral symmetry: L = U i − m U = U Li

L

+ U Ri

R

− m( U L

R

+ UR

L)

:

(1.7)

The original U (1)L × U (1)R chiral symmetry of the massless theory has now broken to a residual U (1)L+R , which is the vectorial symmetry of fermion number conservation. We can see explicitly that the vector current is conserved since the transformation Eq. (1.5) with , = ! is still a symmetry of the Lagrangian Eq. (1.7). The axial current, on the other hand, is no longer conserved: ←

→

9% U % 5 = U 9= 5 + U 5 9=

= −2im U 5

:

(1.8)

The Dirac mass term has spoiled the axial symmetry (, = −!). (iii) Spontaneously massive fermion → Nambu–Goldstone boson Through a sleight of hand, however, we can preserve the full U (1)L × U (1)R chiral symmetry, and still give the fermion a mass! We introduce a complex scalar <eld . with a Yukawa coupling (g) to the fermion. We assume that . transforms under the U (1)L × U (1)R chiral symmetry as . → exp[ − i(, − !)].

(1.9)

that is, . has nonzero charges under both the U (1)L and U (1)R symmetry groups. Then, we write the Lagrangian of the system as L = U Li

L

+ U Ri

R

− g( U L

R.

+ UR

L.

∗

) + L. ;

(1.10)

where L. = |9.|2 − V (|.|) :

(1.11)

Unlike the previous case where we added the fermion mass term and broke the symmetry of the Lagrangian, L remains invariant under the full U (1)L × U (1)R chiral symmetry transformations. The vector current remains the same as in the pure fermion case, but the axial current is now changed to →

←

j%5 = U % 5 + 2i.∗ ( 9 % − 9 % ). :

(1.12)

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243

We can now arrange to have a “spontaneous breaking of the chiral symmetry” to give mass to the fermion. Assume the potential for the <eld . is V (.) = −M 2 |.|2 + 12 1|.|4 :

(1.13)

The vacuum √ built around the <eld con
(1.14)

We can parameterize the “small oscillations” around the vacuum state by writing .=

√1

2

(v + h(x)) exp(i'(x)=f) ;

(1.15)

where '(x) and h(x) are real <elds. Substituting this ansatz into the scalar Lagrangian (1.11) we obtain 1 3 1 4 1 2 2 2 Mh − 1h L. = (9h) − M h − 2 2 8 √ 2M v2 1 2 2 2 + 2 (9') + h (9') + h(9')2 + ; (1.16) 2 2f 2f 1f2 where we have a negative vacuum energy density, or cosmological constant, =−M 4 =21 (of course, we can always add a bare cosmological constant to have any arbitrary vacuum energy we wish). We see that '(x) is a massless <eld (a Nambu–Goldstone mode). It couples only derivatively to other <elds because of the symmetry ' → ' + ). 4 The <eld h(x), on the other hand, has a positive mass-squared of m2 = 2M 2 . The proper normalization of the kinetic term, for ', i.e., (v2 =2f2 )(9')2 , requires that f = v. Again, f is the decay constant of the pion–like object '. The decay constant f is √ always equivalent to the vacuum expectation value (apart from a possible conventional factor like 2). Notice that the mass of h(x) can be formally taken to be arbitrarily large, i.e., by taking the limit M → ∞, and 1 → ∞ we can hold v2 = f2 = 2M 2 =1 <xed. This completely suppresses Iuctuations in the h <eld, and leaves us with a nonlinear model [3]. In this case only the Nambu– Goldstone √ ' <eld is relevant at low energies. In the nonlinear model we can directly parameterize . = (f= 2) exp(i'=f). The axial current then becomes j%5 = U % 5 − 2f9% ' ;

(1.17)

where the factor of 2 in the last term stems from the axial charge 2 of . (Eq. (1.9)). Let us substitute this into the Lagrangian Eq. (1.10) containing the fermions: √ L = U L i L + U R i R + 12 (9')2 − (gf= 2)( U L R ei'=v + U R L e−i'=f ) : (1.18) 4

This is a general feature of a Nambu–Goldstone mode, and implies “Adler decoupling”: any NGB emission amplitude tends to zero as the NGB four-momentum is taken to zero.

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If we expand in powers of '=f we obtain: √ √ L = U i + U i + 12 (9')2 − (gf= 2) U − i(g= 2)' U 5 + · · · :

(1.19)

√ We see that this Lagrangian describes a Dirac fermion of mass m = gf= 2, and a massless √ pseuU doscalar Nambu–Goldstone boson ', which is coupled to i 5 with coupling strength g = 2m=f. This last result is the “unrenormalized Goldberger–Treiman relation” [33]. The Goldberger–Treiman relation holds experimentally in QCD for the axial coupling constant of the pion gA and the nucleon, with m = mN , f = f , and is one of the indications that the pion is a Nambu–Goldstone boson. The Nambu–Goldstone phenomenon is ubiquitous throughout the physical world, including spin-waves, water-waves, and waves on an in
(1.20)

This is gauge invariant in the usual way, since with A% → A% +9% & we can rephase . as . → e−ie& .. Thescalar potential V (.) is as given in Eq. (1.13), hence, . will again develop a constant VEV v= 2=1M . We can parameterize the oscillations around the minimum as in Eq. (1.15) and introduce the new vector potential, B% = A% −

1 9% ' : e

(1.21)

Lagrangian (1.20) in this reparameterized form becomes L. = − 14 F%( F %( + 12 e2 v2 B% B% + 12 (9% h)2 √ 1 3 1 4 1 2 2M 2 2 2 Mh − 1h + e h + h B% B% ; M h − 2 8 2 1

(1.22)

where M is de
C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 R,

with an Abelian Higgs model: L = L + U L (i − eA) L + U R i .

R

− g( U L

R.

+ UR

L.

†

):

245

(1.23)

The Lagrangian is completely invariant under the electromagnetic gauge transformation. Also, the theory is intrinsically “chiral” in that the left-handed fermion has a di,erent gauge charge than the right-handed one. Now we see that, upon writing . as in Eq. (1.15), and performing a <eld rede
L

−mU −

√1

2

gh U

:

(1.24)

Thus we have generated: (1) a dynamical gauge boson mass, ev, and (2) a dynamical fermion √ Dirac mass m = gv= 2. The Dirac mass mixes chiral fermions carrying diQerent gauge charges, and would super
Unfortunately, this simple model is not consistent at quantum loop level, since an axial anomaly occurs in the gauge fermionic current. This can be remedied by, e.g., introducing a second pair of chiral fermions with opposite charges.

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U interactions bind quark anti-quark pairs into a composite 0+ <eld 33. This develops a non-zero 3 U vacuum expectation value 33 ≈ QCD , in analogy to the Higgs mechanism. This, in turn, spontaneously breaks the chiral symmetry SU (2)L × SU (2)R down to SU (2) of isospin. The light quarks then become heavy, developing their “constituent quark mass” of order mNucleon =3. The pions, the lightest pseudoscalar mesons, are the Nambu–Goldstone bosons associated with the spontaneous symmetry breaking and are massless at this level (the pions are not identically massless because of the fundamental quark masses, mu ∼ 5 MeV, md ∼ 10 MeV). The essence of this dynamics is captured in a toy model of QCD chiral dynamics known as the Nambu–Jona-Lasinio (NJL) model [34,35]. The NJL model is essentially a transcription to a particle physics setting of the BCS theory of superconductivity. In Appendix B we give a treatment of the NJL model. If we follow these lines a step further and switch o, the Higgs mechanism of the electroweak interactions, then we would have unbroken electroweak gauge <elds coupled to identically massless U quarks and leptons. However, it is apparent that the QCD-driven condensate 33 = 0 will then spontaneously break the electroweak interactions at a scale of order QCD . The resulting Nambu– Goldstone bosons (the pions) will then be eaten by the gauge <elds to become the longitudinal modes of the W ± and Z bosons. The chiral condensate characterized by a quantity QCD ∼ f would then provide the scale of the W ± and Z masses, i.e., the weak scale in a theory of this kind is given by vweak ∼ f . Because f ≈ 93 MeV is so small compared to vweak ∼ 175 GeV, the familiar hadronic strong interactions cannot be the source of EWSB in nature. However, it is clear that EWSB could well involve a new strong dynamics similar to QCD, with a higher-energy-scale, ∼ vweak , with chiral symmetry breaking, and “pions” that become the longitudinal W ± and Z modes. This kind of hypothetical new dynamics, known as Technicolor, was proposed in 1979 (Section 2). (ii) Naturalness Various scienti
s () → ∞. 6 More introspectively, the scale QCD arises from the explicit scale breaking in QCD that is encoded into the “trace anomaly”, the divergence of the scale current S% : 9% S% = T%% = −

(gs ) %( G G%( ; 2gs

(1.26)

where ( s ) is the QCD- function, arising from quantum loops, and is of order ˝ (we neglect quark masses and, indeed, this has nothing to do with the quarks; the phenomenon happens in pure QCD). The smallness of s at high energies implies that scale invariance is approximately valid there. 6

Perhaps an enterprising string theorist will one day compute s (MPlanck ), obtaining a plausible result such as

s (MPlanck ) ∼ 1=42 thus completely explaining the detailed origin of QCD .

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247

Asymptotic freedom implies that as we descend to lower energy scales, s slowly increases, until the scale breaking becomes large,
This point is actually somewhat more subtle; scale symmetry can in principle act as a custodial symmetry if there are no larger mass scales in the problem; see [36].

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heavy top quark. Here one takes the point of view that there are, indeed, fundamental scalar <elds in nature, and they are governed by the organizing principle of SUSY that mandates their existence. This leads to the MSSM in which all of the Standard Model <elds are placed in N =1 supermultiplets, and are thus associated with superpartners. SUSY and the electroweak symmetry must be broken at similar energy scales to avoid unnaturally
The Higgs boson is then the analogue of the -meson in QCD, which is a very wide state, diVcult to observe experimentally, and can be decoupled in the nonlinear -model limit.

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249

model-building trends such as SUSY, and to re-examine ideas about Iavor physics. Because experiment has played such a key role in guiding the development of these theories, we choose to present the phenomenological analysis in parallel with the theoretical. Each set of experimental issues is introduced at the point in the theoretical story where it has had the greatest intellectual impact. Section 2 explores the development of pure Technicolor theories. As already introduced in this section and further discussed in Section 2.1, pure Technicolor (an asymptotically free gauge theory which spontaneously breaks the chiral symmetries of the new fermions to which it couples) can explain the origins of EWSB and the masses of the W and Z bosons. Section 2.2 discusses the mathematical implementation of these ideas in the minimal two-Iavor model and the resulting spectrum of strongly-coupled techni-hadron resonances. The phenomenology of these resonances and the prospects for discovering new strong dynamics in studies of vector boson scattering at future colliders are also explored. The one-family TC model and its rich phenomenology are the subject of Section 2.3. A more realistic Technicolor model must include a mechanism for transmitting EWSB to the ordinary quarks and leptons, thereby generating their masses and mixing angles. The original suggestion of an Extended Technicolor (ETC) gauge interaction involving both ordinary and techni-fermions alike is the classic physical realization of that mechanism. As discussed in Sections 3.1 and 3.2, the extended interactions can cause the strong Technicolor dynamics to aQect well-studied quantities such as oblique electroweak corrections or the rates of Iavor-changing neutral current processes. Moreover, the extended interactions require more symmetry breaking at higher energy scales, so that the merits of the weak-scale theory are, as with SUSY, entwined with mechanisms operating at higher energies. These issues have had a profound inIuence on model-building. Section 3.3 describes some of the explicit ETC scenarios designed to address questions of Iavor physics, further symmetry breaking, and uni
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dynamics, one is led back to acceptable schemes under the rubric of Topcolor-Assisted Technicolor (TC2). The TC2 models incorporate the best features of the TC and Topcolor ideas in order to explain the full spectrum of fermion masses, while avoiding the classic isospin violation and FCNC dilemmas that plague traditional ETC models. The Topcolor theory, its relationship to TC, and associated phenomenology are the focus of Sections 4.2 and 4.3. Further insights into the dynamics of mass generation have arisen in the context of Top-Seesaw models (Section 4.4), in which the top quark’s large mass arises partly through mixing with strongly-coupled exotic quarks. Most recently, as discussed in Section 4.5, Topcolor is a forerunner of and has a natural setting in latticized or “deconstructed” extra dimensions [41,42]. Topcolor may represent a connection between the phenomenlogy of EWSB and the possible presence of extra-dimensions of space–time at the ∼ TeV scale. All in all, new information about the top quark and new ideas about the structure of space–time have fostered a mini-renaissance in the arena of new strong dynamics and EWSB. 2. Technicolor Motivations underlying Technicolor have been described for the reader in Section 1. We wish to mention a few of the many detailed earlier reviews. The review of Farhi and Susskind [43] and vintage lectures by various authors [44–50] remain useful introductions. There also exists a collection of reprints [51] tracing the early developments. To our knowledge there is no comprehensive review of the “medieval” period of TC, ca. late 1980s to early 1990s. For more recent surveys, the reader should consult the reviews of Lane [52,53] and Chivukula [54–56]. King has also written a more recent review [57] which develops some speci
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fermions are approximately massless, and seek a mechanism to provide only the heavy gauge boson masses. TC was a natural solution to this problem. 2.1.1. The TC ↔ QCD analogy TC is a gauge theory with properties similar to those of QCD. For concreteness, consider a TC gauge group GT = SU (NT ), having NT2 − 1 gauge bosons, called “technigluons.” We introduce identically massless chiral “techniquarks” subject to this new gauge force: QLai and QRai , where a refers to TC and i is a Iavor index. We will assume that the Q’s fall into the fundamental, NT representation of SU (NT ). We further assume that we have NTf Iavors of the Q’s. This then implies that we have an overall global chiral symmetry: SU (NTf )L × SU (NTf )R × [U (1)A ] × U (1)Q (where the U (1)A is broken by the axial anomaly and is thus written in the square brackets [ : : : ]). We will call this the “chiral group” of the TC theory. TC, like QCD, is assumed to be a con
(2.27)

This phenomenon occurs in QCD, in analogy to the Cooper pair condensate in a BCS superconductor, and gives rise to the large nucleon and/or constituent quark masses. The general implication of Eq. (2.27) is the occurrence of a “mass gap”: the techniquarks acquire constituent masses of order m0 ∼ T . Technibaryons composed of NT techniquarks will be heavy with masses of order ∼ NT T . 2 There must also occur NTf − 1 massless Nambu–Goldstone bosons, with a common decay constant FT ∼ T . By analogy, in QCD, if we consider the two Iavors of up and down quarks to be massless, then there is a global chiral symmetry of the Lagrangian of the form SU (2)L ×SU (2)R ×[U (1)A ]×U (1)B , (where “A” stands for axial, and “B” for baryon number). Within an approximation to the chiral dynamics of QCD, known as the “chiral constituent quark model” (e.g., see [70]), based upon the Nambu–Jona-Lasinio (NJL) model, (Appendix B) we can give a description of the dynamical chiral symmetry breaking in QCD or TC (we will refer to this as the NJL model below). In this approximation, we obtain the Georgi–Manohar [71] chiral Lagrangian for constituents quarks and mesons as the low energy solution to the model. In this NJL approximation there is a cut-oQ scale M , which is of order ∼ m , and we can relate f to the dynamically generated “mass gap” of the theory, i.e., the “constituent quark mass” of QCD. If Nf fermion Iavors condense, each having Nc 9

Note that this is not the case in other schemes, such as Topcolor, and a spontaneously broken noncon
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colors, then the quarks will have a common dynamically generated constituent mass m0 and produce a common decay constant, f for the (Nf2 −1) Nambu–Goldstone bosons given by the Pagels–Stokar relation [72] (see Appendix B; in the next section we discuss normalization conventions for f ): f2 =

Nc 2 m ln(M 2 =m20 ) : 42 0

(2.28)

In the NJL approximation we also obtain an explicit formula for the quark condensate bilinear: QU iL QjR = -ij

Nc m0 M 2 : 82

(2.29)

Improvements can be made to the NJL model by softening the four-fermion interaction and treating technigluon exchange in the ladder approximation (see, e.g., [73–75]). Alternatively, lattice gauge theory techniques can be brought to bear upon TC as well (see, e.g., [76,77]). The TC condensate is diagonal in an arbitrary basis of techniquarks, Qi ’s, where the chiral subgroup, SU (NTf )L × SU (NTf )R × U (1)Q , is an exact symmetry (U (1)A is broken by instantons, and the Techni-< is heavy like the < of QCD). The Standard Model gauge interactions, SU (3) × SU (2)L × U (1)Y will be a gauged subgroup of this exact TC chiral subgroup. Indeed, since we want to dynamically break electroweak symmetries, then SU (2)L × U (1)Y must always be a subgroup of the chiral group. In the minimal model described in the next section, QCD is not a subgroup of the chiral group, while in the Farhi–Susskind model, both QCD and electroweak gauge groups are subgroups of the chiral group. ai When the SU (3) × SU (2)L × U (1)Y interactions are turned on, a particular basis for the QL; R has thus been selected (the general models of Q’s contains various SU (3) × SU (2)L × U (1)Y representations). Thus, an “alignment” occurs in the dynamical condensate pairing of QU iL with QRi . In general it is not obvious ab initio that this alignment preserves the exact gauge symmetries (like electromagnetism and QCD; i.e., the electric charge and color generators must commute with the condensate to preserve these symmetries), and breaks yet other symmetries (electroweak) in the desired way. This is one example of the “vacuum alignment” problem [78–85]. In the simplest TC representations the desired vacuum alignment is manifest. 2.1.2. Estimating in TC by rescaling QCD; f , FT , vweak Since TC is based upon an analogy with the dynamics of QCD, we can use QCD as an “analogue computer” to determine, by appropriate rescalings, the properties of the pure TC theory . A convenient set of scaling rules due originally to ‘t Hooft [86–88] (see, furthermore, e.g., [71,89]), characterize the behavior of QCD. These rules have been extensively applied to TC [90]. The main scaling rules are f ∼ Nc QCD ; QU i Qj ∼ -ij Nc 3QCD ; m0 ∼ QCD : (2.30) These rules follow from the NJL approximation with the identi
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Standard Model we can parameterize the Higgs <eld with its VEV as √ √ = 2 + h = 2 v 0 0 H = exp(ia = a =v0 ) : 0

253

(2.31)

This gives the kinetic terms for the a (and h0 ) the proper canonical normalizations in the limit of switching oQ the gauge <elds. From the Higgs boson’s kinetic terms we extract, where the electroweak covariant derivative D% is de
F ≈ 93 MeV ;

(2.33)

a where j%a5 = U % 5 =2 where = (u; d) in QCD (Note: another de
U = exp(ia = a =f);

L=

f2 Tr(9% U † 9% U ) : 4

(2.34)

Then the normalization is f = f = 93 MeV, which can be seen by working out the axial current, j%5 = -L=-9% a and comparing with Eq. (2.33). We will similarly de
FT ˙ vweak :

(2.35)

Including electroweak gauge interactions the techniquark kinetic terms take the form QL iDQL + QR iDQR →

FT2 Tr((D% U )† (D% U )) ; 4

(2.36)

where D% is de
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Hence, comparing Eq. (2.32) to Eq. (2.37) we see that the Higgs VEV v0 = FT when we have a single doublet of technipions. If ND doublets carry weak √ charges then Eq. (2.36) contains ND terms. FT remains the same, but the weak scale becomes v0 = ND FT . 10 Consider a TC gauge group SU (NT ) with ND electroweak left-handed doublets and 2ND singlets of right-handed techniquarks, each in the NT representation. The strong SU (NT ) gauge group will form a chiral condensate pairing the left-handed fermions with the right-handed fermions. This produces (2ND )2 − 1 Nambu–Goldstone bosons (technipions, T , and the singlet <T ) each with decay constant FT . Hence, we can estimate FT from the QCD analogue f using the scaling rules: T T NT ND NT f ; v0 = ND FT ∼ f : (2.38) FT ∼ 3 QCD 3 QCD As stated above, v02 receives contributions from ND copies of the electroweak condensate. Now, in the above example of scaling, we have taken the point of view that T is <xed and, e.g., √ v0 varies as ∼ NT ND as we vary NT and ND . This is an unnatural way to de
10

If we could switch oQ nature’s EWSB, the preceding discussion indicates precisely how QCD itself would then break the electroweak interactions. Indeed, (see e.g. [52]) if there is no EWSB (no Higgs boson), then the up and down quarks are identically massless. Then we would have the QCD chiral condensate, generating constituent quark masses for up and down of order 300 MeV, and massless composite pions in the absence of electroweak gauge interactions. In the presence of electroweak gauge interactions the pions are mathematically “eaten” to become WL and ZL . Now, however the W and Z masses given by MW = f MW =v0 ∼ 29 MeV and MZ = f MZ =v0 ∼ 33 MeV. The longitudinal W and Z would thus be the ordinary ’s of QCD. Thus, QCD misses the observed masses by ∼ 4 orders of magnitude, but it gets the ratio of MW =MZ correct!

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255

As “a” is the TC index, we have NT TC copies of these objects. Anomalies involving Y are absent because we have introduced a vector-like pair (TR ; BR ), each element with opposite Y (the Witten global SU (2) anomaly [91] vanishes provided NT is even for any ND ). We can readily generalize the model to include arbitrary Iavors, ND ¿ 1 doublets, of the same color-singlet technifermions. With these assignments the techniquarks have electric charges as de
T (MGUT ) = 3 (MGUT ) ;

(2.41)

then evolving down by the renormalization group, we obtain for the TC scale T to one-loop precision:

T 2(b0 − b0 ) ; (2.42) = exp QCD b0 b0 3 (MGUT ) where, 2 11NT 4 (2.43) nf and b0 = − ND 3 3 3 b0 and b0 are the one-loop -function coeVcients of QCD and TC, respectively. Putting in some “typical” numbers, we
2.2.2. Spectroscopy of the minimal model (i) Techniquarks The spectrum of the minimal model follows QCD as well. The mass gap of the theory can be estimated by scaling from QCD. √ From √ Eq. (2.39) the techniquarks will acquire dynamical constituent masses of order: mTQ ∼ m0 v0 3=f NT ND where m0 ∼ mN =3 ∼ 300 MeV is the constituent quark mass in QCD. This gives mTQ ∼ 690 GeV for the minimal model with NT = 4, ND = 1. Hence, there √ will be a spectrum of “baryons” composed of QQQQ with a mass scale of order ∼ 3= ND TeV. At the pre-ETC level the lightest of these objects is stable, and cosmologically undesirable [64]. However, in the presence of the requisite ETC interactions there will be Q → q + X , Q → ‘ + X

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transitions and the baryons will become unstable to decay into high multiplicity
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where the parameter 1 describes the ETC coupling of
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C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 τa

0

0

τa

T1 B1 T2 B2

+ - 0 - 0 π+, π , π = WL, WL, Z L

0

0

τ

0

- τa

τ

a

I

0

0

-I

τ

a

a

0

iτ

0

-i τ a

0

0

I

0

iI

I

0

-iI

0

a

9 PNGB’s

3 Techniaxions

Fig. 1. The (2ND )2 − 1 Nambu–Goldstone bosons corresponding to an ND = 2 extension of the minimal model. The
are the electrically charged technipions—hence, it is easier to examine the properties under U (1)EM . In contrast, the techniaxions are neutral under U (1)Y , rendering them sterile under all of the gauge interactions. To estimate the masses of the charged PNGBs, we again use QCD as an analogue computer, and rescale the ± − 0 mass-squared diQerence. The mass-squared’s of the non-neutral weak-charged PNGBs are therefore estimated to be of order m2T ∼ 2 2T or mT ∼ 6 GeV [97]. There are further small corrections when we switch on U (1)Y . Unfortunately, charged scalars with such low masses are ruled out experimentally (see Section 3), another diVculty for the minimal model. All electrically neutral PNGBs remain massless at this level of model-building. In particularly, the techniaxions remain perturbatively massless, since they correspond to a residual exact global symmetry (spontaneously broken). Their associated axial currents have electroweak anomalies like the 0 , but no QCD anomaly, and this can lead to a miniscule electroweak instanton source of mass, ˙ exp(−42 =g22 )v0 . These objects behave like axions, with decay constant FT . They would be problematic, since axion-like objects have restricted FT ’s, typically ¿ 108 –1010 GeV by the usual astrophysical arguments. As we will see in Section 3, these problems are ameliorated, in part, by the eQects of ETC, which provide a stronger source of gauge mass for technipions. Incidently, there is expected to be a ,-angle in TC, and it is ultimately interesting as a potential novel source of CP-violation (see e.g., [97]). (iii) Vector mesons In TC there will generally occur isovector and isosinglet s-wave vector mesons, the analogues of 0 0

(770) and !(782) in QCD, which we denote: ± T ; T and !T . The vector mesons are particularly important phenomenologically, because of their decays to weak gauge bosons and technipions (e.g.,

T → WW is the direct analogue of the QCD process → ). The vector mesons provide potentially visible resonance structures in processes like pp or e+ e− → W + W − and, more generally, make large contributions to technipion pair production. The masses of vector mesons can be estimated

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259

by TC scaling, m√

∼ m!TC ∼ m (FT =f ) ∼ m (v0 =f ) 3=NT ND which yields the approximate value of m ; !TC ∼ 1:8= ND TeV for NT = 4. Let us follow the conventional discussion of vector mesons in QCD by introducing the dimensionless phenomenological “decay constants” f TC and f!TC : aT |j%b |0 = H% -ab

m2 TC ; f TC

!T |j%0 |0 = H%

m2!TC ; f TC

(2.45)

√ U % Q= 2, j%0; a = Q U % (= a =2)Q, and Q is a techniquark doublet. It is diVcult to extract where j%0 = Q f! from QCD due to ! − ' mixing, so one typically assumes “nonet” symmetry, f = f! . We then determine the decay constants from the partial width, E( 0 → e+ e− ) = 4 m =3f 2 ; this implies f ≈ f! ≈ 5:0. We must determine how f and f! undergo TC scaling. Note that the current matrix elements of Eq. (2.45) involve a TC singlet combination of techniquarks, and a normalized initial state, and 2 might therefore be expected to scale from QCD as ∼ N =3( = T T QCD ) . However, for <xed v0 we 2 2 2 2 write this as ∼ (v0 =f ) 3=NT ND . so the amplitude m TC =f TC must scale as ∼ 3=NT ND2 . Since m TC ∼ 3=NT ND we see that [98] (2.46) f TC ∼ 3=NT f ∼ 4:3 for NT = 4 : This result makes intuitive sense only if one keeps in mind that, in TC scaling, we are holding v0 <xed! The decay modes of the vector mesons have been considered in the literature [98–101]. Some can be treated by scaling from the principle QCD decay modes ± → ± 0 , 0 → 20 , 0 → + − , and ! → + 0 − , ! → 0 . Note that the ! decay modes are associated with anomalies in a chiral Lagrangian description, and have relatively tricky scaling properties [98]. In TC, by invoking the “equivalent Nambu–Goldstone boson rule” for the longitudinal gauge bosons, one
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Table 1 Estimated properties of lowest-lying (pseudo-) scalar and (axial-) vector mesons in the minimal TC model with a single electroweak doublet of techniquarks ND = 1, NT = 4, and q = 3=NT = 0:86 State

I (J PC )

Mass (TeV)

Decay width (GeV)

T± ; T0 → WL± ; ZL <0T a 0

± T , T 0 !T 0 a± 0T ; a0T 0 f0T ∼ T 0 a± 1T , a1T f1T

1(0−+ ) 0(0−+ ) 1(1− − ) 0(1− − ) 1(0++ ) 0(0++ ) 1(1++ ) 0(1++ )

MW ; MZ (eaten) ∼ rm< ∼ 0:4 → 0:8 qm =s ∼ 1:2 ∼ qm! =s ∼ 1:2 rma0 ∼ 1:5 qf0 =s ∼ 2 qma1 =s ∼ 2 qmf1 =s ∼ 2

EW ; EZ Et tU ∼ 8:0 − 64:0, Egg ∼ 0:3 − 3:0 E T (WW ) ∼ E ()=sq2 ∼ 350 E!T (WWZ) ∼ E! ()=sq2 ∼ 80
We take r ≡ T =QCD = 1:5 × 103 , and QCD = 200 MeV, and s ≡ f =FT = 5:7 × 10−4 , where f = 100 MeV, FT = 175 GeV. The combination r 3 s2 = 1:1 × 103 frequently occurs. a These are estimates from the discussion of [104].

0

0

E(!T → Z Z ) ∼ +

−

E(!T → W W ) ∼

3 NT

1=2 3 NT

1:1=

1=2

ND GeV ;

5:2=

ND GeV :

(2.48)

(iv) Higher p-wave resonances The parity partners of the T and !T are the p-wave axial vector mesons, a1T and f1T . If the chiral symmetry breaking of QCD or TC were somehow switched oQ 12 states and their parity partners would be degenerate. Given the presence of spontaneous chiral symmetry breaking, we must estimate the masses of the√axial vector mesons by scaling from QCD: ma1T ≈ mf1T ∼ ma1 ;f1 (1260)(v0 =f ) 3=NT ND ∼ 2:9= ND TeV. The spectrum should also include p-wave parity partners of the T and
There really is no conceptual limit of the true theory that can do this. It would be analogous to taking the coupling constant on the NJL model to be sub-critical; see [102].

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261

2.2.3. Non-resonant production and longitudinal gauge boson scattering Since the longitudinal W and Z are technipions, the minimal TC model predicts that high energy WL − WL , ZL − ZL or WL − ZL scattering will be a strong-interaction phenomenon. Studying the pair-production and scattering of the longitudinal W and Z states thus provides a potential window on new strong dynamics. As Nambu–Goldstone bosons, the longitudinal W and Z are described by a nonlinear -model chiral Lagrangians, [105–116]. This is called “the equivalence theorem”, [117–133], and often this is viewed as an abstract approach, without speci
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Table 2 Number of years at LHC with annual luminosity 100 fb−1 required for a 99% con<dence level signal Model Channel

SM

Scal

O(2N )

ZZ(4‘) ZZ(2() W +W − W ±Z W ±W ±

1.0 0.5 0.75

2.5 0.75 1.5

3.2 1.0 2.5

4.5

3.0

4.2

V1.0

V2.5

CG

LET-K

Dly-K

3.7 8.5 7.5 1.5

4.2

3.5 9.5

4.0

5.7

1.5

1.2

1.2

2.2

100

W± Z → 3l 1ν

10 1 0.1 0.01

(a)

0

1000

2000 MT

3000

Events / (50 GeV 100 fb-1)

Events / (50 GeV 100 fb-1)

The models considered are: the Standard Model (SM), strongly-coupled models with new scalar resonances (Scal, O(2N)), strongly-coupled models with new vector resonances of mass 1 TeV (V1.0) or 2:5 TeV (V2.5), and strongly-coupled models with non-resonant scattering following the low-energy theorems (CG, LET-K, Dly-K). From Ref. [134].

(b)

100

W± Z → 3l 1ν

10 1 0.1 0.01 0

1000

2000

3000

MT

Fig. 2. Event yields at the LHC for T → W ± Z 0 → ‘± (‘ ‘+ ‘− for M T = 1:0, 2:5 TeV; from Ref. [134]. A conventional techni- resonance of mass much above 1 TeV would be invisible in the channel T → W ± Z 0 → ‘± (‘ ‘+ ‘− .

boson, H → W + W − → l(jj and H → ZZ → l+ l− jj, can provide statistically signi
C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402

d σ /dM [nb/(GeV/c2)]

10

−6

263

+ − pp −> W W + anything

−7

10

100 TeV 40 TeV

−8

10

20 TeV

|y| < 1.5 −9

10

1.4

1.6 +

1.8

2.0

-

W W mass (TeV/c2) Fig. 3. Vector meson dominance production of techni- with subsequent decay to W + Z in pp collider with center-of-mass energies, 20, 40 and 100 TeV (from EHLQ [157]).

have been discussed extensively in the literature, beginning with EHLQ [101,103], and [122,98,149– 156]. Examples of the calculated cross-sections for T production and decay at an LHC or VLHC are given in Figs. 3 and 4. Detailed search strategies and limits for the “low-scale” variants of the minimal model are discussed in Section 3.5. Absent a direct coupling between ordinary and techni-fermions such as ETC can provide (Section 3), how are the techni-vector mesons of the pure Minimal TC model to be produced in qqU or e+ e− annihilation? The answer is that vector meson dominance (VMD) enables the techni-vector mesons to couple to currents of ordinary fermions. Most relevant to a discussion of T production ± are vector dominance mixing of the 0T with and Z and the ± T with W , which we discuss below. The production and detection of the techni-! is considered in [98]. Let us brieIy review the theory of VMD. Consider a schematic eQective Lagrangian in which we introduce the photon, A% together with a single neutral vector meson 13 % : % ; L = − 14 (FA%( )2 − 14 (F %( )2 − 12 HFA%( F %( + 12 m2 ( % )2 − eA% J % − N % Jhad

(2.49)

% is a vector where J % is the ordinary electromagnetic current, and we de
13

This can be directly generalized to electroweak gauge <elds and an isotriplet a , or to gluons and a color octet a8 , but see [158,159].

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C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 −6

d σ /dM [nb/(GeV/c 2 )]

10

+ − pp −> ZW + ZW + anything

−7

10

100 TeV

40 TeV

−8

10

20 TeV

|y| < 1.5

−9

10

1.4

1.6

1.8

2.0

WZ mass (TeV/c2) Fig. 4. Vector meson dominance production of techni- with subsequent decay to W ± Z in pp collider with center-of-mass energies, 20, 40 and 100 TeV (from EHLQ [157]).

The H term represents mixing between the photon and , and can be viewed as arising from the ( nonzero amplitude 0|TJ % (0)Jhad (x)|0 ; we will work to order H. Note that the vector meson, , can be viewed as a gauge <eld that has acquired mass through spontaneous symmetry breaking [160] (Indeed, Bando, Kugo, Yamawaki and others [161–164] have argued that vector meson eQective Lagrangians always contain a hidden local symmetry). This is why we choose the kinetic term to be in the form of the photon kinetic term, and it implies that we are always free to choose a gauge, such as 9% % = 0. Upon integrating by parts, we can rewrite the H term as HA( 9% F %( . Using equations of motion for the to order H we obtain: % ( L = − 14 (FA%( )2 − 14 (F %( )2 + Hm2 A( ( + HNA( Jhad + 12 m2 ( % )2 − eA% J % − N % Jhad :

(2.50)

The
14

(2.51)

This does not violate gauge invariance. While a shift A% → A% + 9% , leads to Hm2 9( , ( , we can integrate by parts −Hm2 ,9( ( , and the gauge invariance of the allows this term to be set to zero. That is, % behaves like a conserved current if it is a gauge <eld with hidden local symmetry [164].)

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265

( This removes the mass mixing term, and the HA( Jhad term, but leads to an induced H % J % term. Thus, we can view the as having an induced direct coupling to the full electromagnetic current of strength H! Alternatively, upon integrating by parts, we could have written the H term as H ( 9% FA%( , and using equations of motion for the to order H we have He ( J ( = e2 ( J ( =f . Thus, we can view the eQect of the H term as directly inducing the coupling of the to any electromagnetic current with strength e2 =f , e.g., the will couple directly to the electron’s electromagnetic current. While this is a small coupling, the is generally a narrow state. On-resonance the production rate can be substantial. An equivalent non-Lagrangian description of this treats the H term as a mixing eQect in the propagator of the photon. The propagator becomes a matrix, allowing the photon to mix with the

and couple directly to Jhad . The propagator connecting an electromagnetic current to the hadronic current becomes:

−i −i × −iHq2 × 2 : 2 q q − m2

(2.52)

In the context of VMD, we can summarize our expectations for the minimal model with ND = 1. We have already estimated the techni- mass to be of order ∼ 1:8 TeV in this case. The dominant expected production and decay modes of the T are then ffU → ( ; Z 0 ) → 0T → W + W −

(2.53)

± 0 ffU → (W ± ) → ± T → W Z :

(2.54)

and

The annihilation of a fermion and anti-fermion a T and its subsequent decay into technipion + gauge boson, or technipions can also be signi
Tandean [93] has also considered the signature and detectability of the
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2.3. Farhi–Susskind model 2.3.1. Structure The minimal model is neither a unique prescription for the construction of a TC theory, nor is it likely to contain suVcient richness to ultimately allow the generation of the observed range of fermion masses and mixing angles. We anticipate that a more complete model would need to include both a “quark” sector of color-triplet techniquarks, and a “leptonic” sector of color-singlet technifermions (like those in the minimal model). These states could ultimately act under extended TC interactions 17 to give mass terms to the ordinary quarks and leptons. Toward this end, we turn now to describing the Farhi–Susskind TC model [165] which contains a much richer spectrum of technifermions imitating the anomaly free representations of the Standard Model. The Farhi–Susskind model extends the Iavor content of the minimal model to imitate one full generation of quarks and leptons with the usual anomaly free isospin and Y assignments: i T N QL = Y = y; Y = −3y ; B L E L QR = (TRi ; BRi NR ; LR );

Y = (y + 1; y − 1; −3y + 1; −3y − 1) ;

(2.55)

where the SU (3) color index i takes the values i = 1; 2; 3 for the techniquarks. This is an anomaly free representation for any choice of the parameter y. For the particular standard choice of y = 1=3, the techniquarks and technileptons have electric charges identical to those of the quarks and leptons. In the present model the techniquarks and technileptons couple to the full SU (3)×SU (2)L ×U (1)Y gauge group in the usual way. Now we postulate that each Q multiplet carries, in addition, an SU (NTC ) quantum number in the NTC representation of the strong TC gauge group. Note that the existence of a right-handed technineutrino, NR , is required to provide anomaly cancellation for the TC gauge interaction. The SU (NTC ) is essentially a gauged horizontal generation symmetry. Unlike the minimal model with ND =1, in which all three NGBs are absorbed into the longitudinal modes of the electroweak gauge bosons, the Farhi–Susskind model has a low energy spectrum containing numerous pseudo-Nambu–Goldstone bosons (PNGBs). Their quantum numbers may be ascertained by observing that, in the limit of vanishing SU (3) × SU (2)L × U (1) couplings, there is an SU (8)L × SU (8)R × U (1)A × U (1) global chiral group. The full Standard Model SU (3) × SU (2)L × U (1)Y interactions are a gauged subgroup of this chiral group. At the scale TC , the TC gauge coupling is strong, and causes a degenerate chiral condensate to form: TU Li TRi = BU Li BRi = NU L NR = EU L ER ∼ 3TC ;

(2.56)

where i is a (unsummed) color index ranging from i = 1; 2; 3. The chiral group is thus broken spontaneously to an approximate SU (8) × U (1) vectorial symmetry, producing 63 + 1 NBBs. From the previous remarks about condensates, we see that there exist four composite electroweak doublets. This is similar to the structure of a 4-Higgs-doublet model in which each Higgs boson gets a common VEV FT . The electroweak scale is thus related to the common VEVs of the four Higgs bosons, FT , as v02 = 4FT2 , and thus FT = 123 GeV. One combination of the NGBs will become 17

See Section 3.

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Table 3 Properties of scalar states in the Farhi–Susskind TC model following [103,101]; see also [90] State

I (J P ), color, [Q]

Mass (GeV)

T− ∼ (TU i Bi + NU E) U T0 ∼ (TU i T i − BU i Bi + NU N − EE) U
1[0− ; 1− ] 0[ − 1] 1[0− ; 1− ] 0 [0] 1[0− ; 1− ] 0 [0] 1[0− ; 1− ] 0 [1] 1[0− ; 1− ] 0 [0] 0[0− ; 1− ]1[0] 1(0− ) 3 [5=3] 1(0− ) 3 [2=3] 1(0− ) 3 [ − 1=3] 0(0− ) 3 [2=3] 1(0− ) 8 [1] 1(0− ) 8 [0] 0(0− ) 8 [0]

MW MZ ∼ 103 ∼ 100(4=NTC )1=2 ∼ 100[ETC] ∼ 100[ETC] ∼ 160 (4=NTC )1=2 ∼ 160 (4=NTC )1=2 ∼ 160 (4=NTC )1=2 ∼ 160 (4=NTC )1=2 ∼ 240 (4=NTC )1=2 ∼ 240 (4=NTC )1=2 ∼ 240 (4=NTC )1=2

To each listed spin-0 state there is a corresponding s-wave spin-1 analogue; there will also be p-wave analogue resonance states. Nonself-conjugate states have corresponding (unlisted) antiparticles. Quoted masses are crude estimates, quoted in GeV; their theoretical values are very model dependent, modulo walking ETC, etc. For analogue vector and resonance masses see the discussion of the text.

the longitudinal W and Z, while the orthogonal states remain in the spectrum, as we will describe in the next subsection. 2.3.2. Spectroscopy (i) Color {1}, {3}, and {8}, pseudo-Nambu–Goldstone bosons Let us examine the content of the low lying (8 × 8) PNGBs of the Farhi–Susskind model. As might be expected, the enhanced variety of technifermions yields a larger selection of PNGB states. Their properties are summarized in Table 3. We will begin with the eight color-singlet states. By symmetry, three linear combinations are identically massless and become the longitudinal W ± and Z 0 : TU i Bi + NU E ∼ − ;

U ∼ + ; BU i T i + EN

U ∼ L0 : TU i T i − BU i Bi + NU N − EE

(2.57)

There remain 5 orthogonal color singlet objects, two with non-zero electric charge (we follow the nomenclature and normalization conventions of Eichten, HinchliQe, Lane and Quigg (EHLQ) [103,101]): TU i Bi − 3NU E ∼ P − ;

U ∼ P+ BU i T i − 3EN

(2.58)

and three which are electrically neutral: U ∼ P0; TU i T i − BU i Bi − 3(NU N − EE)

U ∼ P0 TU i T i + BU i Bi − 3(NU N + EE)

(2.59)

and U ∼ <T : TU i T i + BU i Bi + NU N + EE

(2.60)

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The PNGB which is neutral under all gauge interactions receive mass via instantons. The <T , receives mass from the instantons of TC, and is expected to be heavy, as in the case√of the Weinberg–Susskind model. From the discussion of Section 2.2.2(ii) we estimate m< TC ∼ ( 6=NT ) × 3=NT ND (v0 =f )m< ∼ 700 GeV for NT = 4 (where ND = 4 in the Farhi–Susskind model). The P 0 receives a mass only of order 1 GeV from QCD instantons, but will receive a larger contribution from ETC eQects (see Section 3). The P 0 , likewise, receives its mass from ETC. The PNGBs with electroweak gauge charges, but no color, P0Q , receive masses from the gauge interactions in analogy to the electromagnetic mass splitting of the ordinary + and 0 of QCD. In QCD we have -m2 = m2+ − m20 ≈ (35 MeV)2 . In the chiral limit (current masses mu = md = 0) we have m20 = 0, and thus -m2 roughly reIects the electromagnetic mass contribution of the ± . Scaling from this, one
(2.61)

M (P8 ) ∼ 240(4=NTC )1=2 GeV :

(2.62)

(here we use 3 (FT ) = 0:1, (FT ) = 1=128, C2 (3) = 4=3, C2 (8) = 3). More generally, the various electroweak and strong contributions to the mass of a colored PNGB are added in quadrature to form the full mass, m2 = m2c + m2EW . In a more complete model, there can also be Extended TC contributions to the masses, but these are expected to be smaller than the QCD masses given above. The reader is advised to consider the spectrum of states, and reasonable decay and production estimates (which follow), but not take literally the model estimates of masses at this stage. Any TC model we review from the 1980s is, at best, incomplete, and can only serve as a guide to what may be contained in more modern reincarnations of the models. Moreover, in subsequent sections on ETC, Walking TC and hybrid models like TC2 many mechanisms will surface that can rearrange the masses of states in these models. Color-triplet (leptoquark) technipions decay via Extended TC as P3 → q‘U (without the standard choice of y=1=3 these objects would be stable). The color octet technipions decay into P8 → qq+· U ··
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269

Their masses follow from the estimate in Section 2.2.2(iii) if we set ND = 4. Neglecting the QCD corrections, which are expected to be less than ∼ 15%, we
(2.63)

Note that there now exist color-octet V8 states which have the quantum numbers of the gluon, and act like a multiplet of heavy degenerate gluons. 18 These will exhibit vector-dominance-like mixing with gluons in processes like qq U → G → V8 → PA P B . There are also, as in the minimal model, p-wave parity partners of the PNGBs (the techni-a0 ’s and techni-f0 ’s) and parity partners of the vector mesons, the axial-vector mesons a1T ’s and f1T . Following the discussion of Section 2.4, we
(2.64)

M (a0T ; f0T ) ∼ 1300(4=NTC )1=2 GeV :

(2.65)

Generally speaking, the parity-partner states form identical representations of the symmetry groups and have identical charges, but are signi
s2 N 2 m3P 0 E(P → gg) = 3 ; 6 4 F2 2 3 N mP

2 0 E(P → ) = : 3 27 4 F2 U q)) U = E(P 0 → ll(q

(2.66) (2.67) (2.68)

If mP0 ¡ mt =2, then the best hope of
Similar objects will crop up as gauge particles (colorons) in Topcolor models (Section 4.2) or as KK modes in extra-dimensional models (Section 4.6).

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10

1 100

dσ/dy (y=0) (nb)

40

10

0

20 10

10

−1 2 +

p p −> P 0 + anything 10

−2 15

25

Mass Fig. 5. DiQerential cross section for production of P energy in TeV (reproduced from EHLQ [157]).

0

35

45 (GeV/c2)

55

at y = 0 in pp and ppU collisions with indicated center-of-mass

Fig. 6. Projected gg → P 0 → signal rate S and irreducible background rate for an integrated luminosity of 2 fb−1 at Tevatron Run II; from [170].

Recently, Casalbuoni et al. [170] (see also [171]) have looked in detail at the possibility of
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271

(ii) Color-triplet PNGBs: the P3 leptoquarks As we have seen, the spectrum of the Farhi–Susskind TC model includes color-triplet technipions, with leptoquark quantum numbers [43,101,156]. These objects, P3 ’s, would be predominantly resonantly pair-produced through a color-octet techni- (V8 ) coupled either to gluons, or through octet vector dominance mixing (VMD) with the ordinary gluon. The latter case assures a signal in qqU annihilation provided the V8 can be excited. P3 ’s decay (via ETC [172,173]) preferentially to third-generation quarks and leptons. At leading order, the leptoquark pair production cross section depends only on the masses (MV8 ; MP3 ) and the V8 decay width (EV8 ). The latter depends, in turn, on masses, on the size of the TC group NTC and on the mass-splitting between the color-octet and color-triplet technipions (ZM ). While we would expect that V8 VMD with the gluon would provide the largest contribution to the hadronic production of colored PNGBs, provided the V8 pole is within reach of the machine energy, at this writing there is no study of this contribution and its eQects. EHLQ [101], and more recently Skiba [174], have studied of hadron collider signatures of colored PNGBs, without the V8 resonant enhancement. Possible processes in which colored PNGBs can be produced in hadron colliders are: (i) Gluon–gluon and quark–anti-quark annihilations as sources of PNGB pair production; (ii) Quark–gluon fusion producing single PNGBs; (iii) anomalous couplings to two gluons producing single PNGBs. Direct PNGB couplings to fermions (through ETC eQects) are typically too small to give signi

(2.69)

3 kd 2 2 − (1 − z ) (1 − 2V + 2V 2 ) ; d 32

(2.70)

where kd is the “Dynkin index” of the d-dimensional representation (k3 = 12 ; k8 = 3), z the cosine of parton scattering angle in the center of mass system, V =1−

1 − 2 1 − 2 z 2

and

2 = 1 −

4m2P ; sˆ

sˆ and tˆ are Mandelstam variables at the parton level. Using these formulae and parton distributions from Ref. [175] (set 1), Skiba obtains [174] production rates for leptoquarks, P3 , and octets, P8 , which agree with the results of Ref. [101] (see also Hewett et al. [176]). As the cross-section for pair production of leptoquarks at LHC energies is sizeable (see Fig. 7), the LHC can potentially observe leptoquarks with masses up to approximately ∼ 1 TeV. (iii) Color-octet PNGBs, P8 , and vector mesons T 8 Due to color factors, the production cross sections for color octet P8 ’s are an order of magnitude larger than for leptoquarks, as a comparison of Figs. 7 and 8 reveals. The detection of P8 ’s is,

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104 pp -> TT

σ (pp->TT)[pb]

103 102 101 100

16

10-1

4 1.8 200

400 600 mT [GeV]

800

1000

Fig. 7. The cross section for pair production of leptoquarks in pp collisions, for [174]; T = P3 in Skiba’s notation).

√

s = 1:8, 4 and 16 TeV (from Skiba

105

σ (pp->θθ)[pb]

10

pp -> θθ

4

103

16 TeV

102 4 TeV

101

1.8 TeV

100 100

200

300 mθ [GeV]

400

500

Fig. 8. The cross section for pair production of color-octet PNGBs in pp collisions, for Skiba [174]; , = P8 in Skiba’s notation).

√

s = 1:8, 4 and 16 TeV (from

on the other hand, far more diVcult. P8 ’s typically decay into two hadronic jets, so that pair-produced P8 ’s yield four-jet
273

4-jet cross-section (pb/25GeV)

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Mbal (GeV)

Fig. 9. Four-jet rate d=dmbal at a 17 TeV collider with pTmin of 100 GeV. QCD background (hatched) and 240 GeV technipion signal are shown. No resolution eQects included (from Ref. [177]).

background [179–184] and calculating the signal from heavy particle decays. Isolation and centrality cuts must be applied to ensure all jets are detectable. Then appropriate kinematic variables must be chosen to make the signal stand out cleanly above background. The following strategy allows one to pull a P8 signal out of the large QCD background [177], as illustrated in Fig. 9. For each four-jet event, consider all possible pairwise partitions of the jets. Choose the partition for which the two pairs are closest in invariant mass, and de
E(P80 → qq) U =

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Γ (GeV)

274

M P (GeV) U (dashed line), tUt (dotted line) and Z-gluon Fig. 10. Partial widths for the decay of P80 into gluon–gluon (solid line), bb (dot-dashed line) in the Farhi–Susskind model [169].

dσ/dy|y=0 (fb)

108

107

106

105

100

200

300

400

500

600

700

800

900

1000

MP,ρ (GeV)

Fig. 11. DiQerential cross section at y = 0 for single production of 0T 8 (solid line) and P80 (dashed line) at the LHC in femtobarns [169].

E(P80

→ g ) =

Negs 4F

2

m3P : 576

(2.71)

Fig. 10 gives the decay widths of the P80 in the Farhi–Susskind model. Fig. 11 illustrates gluon fusion production at the LHC and we see that the rate of single P80 production is high at a hadron collider. However, the signal in the primary decay channels (gluon or b-quark pairs) is swamped by QCD background. If the PNGB mass is above the t tU threshold, the decay mode P → t tU becomes

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275

dominant and may alter the standard QCD value of the t tU cross section and t tU spin correlations. The decay channel P → g [186,185] holds some promise for the models of “low-scale” TC discussed in Section 3.5; however the PNGBs of the Farhi–Susskind TC model are not likely to be visible in this mode. The 0T 8 coupling to gg and qqU can be estimated by assuming that it mixes with the gluon under a generalized vector meson dominance (VMD). VMD applies to the gg → 0T 8 process as well as the qqU → 0T 8 process [159]. The partial widths relevant for 0T 8 production at hadron colliders in the narrow width approximation are [169]

E( 0T 8 → gg) =

E( 0T 8 → qq) U =

s2 m ; 2

(2.72)

5 s2 m : 6

(2.73)

Fig. 11 compares the single production cross section of the 0T 8 and P80 at the LHC. If kinematically allowed the dominant decay mode of 08T is into two colored PNGB [169]:

E( 08T → P8 P8 ) = T m (1 − 4m2P =m2 )3=2 : (2.74) 4

In this case the 0T 8 contributes strongly to the cross section for color-octet PNGB pair production discussed in above and improves the signal. If the T 8 are light, e.g. in low-scale models (Section 3.5), there can be a sizeable cross section for their pair production. N\aively, well above the pair-production threshold one expects [169]: (pp → T 8 + X ) 1 1 2 : (pp → T 8 T 8 + X ) g T 40

(2.75)

Note that all types of colored vector resonances can be pair-produced, whereas the isosinglet 0T 8 dominates single production. Hence, pair production, unlike single production, can result in interesting decays to longitudinal electroweak gauge bosons. One may also expect spectacular 8-jet events whose kinematics distinguish them from the QCD background, as in the case of color-octet PNGB pair production. The T 8 can signi
C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402

dσ/dM (nb/GeV)

276

M (GeV)

Fig. 12. The ZZ diQerential cross section in nb/GeV vs. MZZ in a toy O(N ) scalar-model with three color-octet PNGBs (upper curve), and the continuum qqU annihilation background (lower curve). A pseudo-rapidity cut |<| ¡ 2:5 is imposed on the
(v) Rescattering An alternative way of
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277

Fig. 13. Sensitivity of a 500 GeV LC with 80 fb−1 of data to M T via the W -boson form factor; from Ref. [193]. The predicted values for a 3-TeV T or non-resonant scattering (LET) lie within the 95% c.l. curve for the prediction of the Standard Model; these cases cannot, then be reliably distinguished. Lighter T give predictions which diQer signi
2.3.4. Production and detection at e+ e− colliders (i) Pair-production of EW bosons The s-channel process e+ e− → WW is an eQective probe of strong electroweak symmetry breaking, especially for physics with a vector resonance [134,192–194]. A 500 GeV linear collider with only 80 fb−1 of integrated luminosity would already be sensitive to radiative corrections induced by vector resonances with masses up to about 2 TeV, but would not be able to observe oQ-resonance contributions to WW production (see Fig. 13). With 500 fb−1 of integrated luminosity, even the non-resonant or low-energy-theorem contributions would become distinguishable from Standard Model expectations, as illustrated in Fig. 14. A higher-energy e+ e− collider would have even greater search √ potential. With s = 1:5 TeV and an integrated luminosity of 190 fb−1 , it should be possible to distinguish the radiative eQects of a very heavy techni- or a non-resonant amplitude from those of the standard model with a light Higgs boson; the 4 TeV (6 TeV) techni- corresponds to a 6:5 (4:8) signal. At a slightly higher integrated luminosity of 225 fb−1 , it would be possible to obtain 7:1, 5:3 and 5:0 signals for a 4 TeV techni- , a 6 TeV techni- , and non-resonant contributions, respectively. The WW fusion processes are complementary to the s-channel W + W − mode since they involve √ more spin–isospin channels (the I =0 and I =2 channels). For an e+ e− collider at s=1:5 TeV with 190 fb−1 of data, the W + W − =ZZ event ratio can be a sensitive probe of a strongly-interacting electroweak sector [134,195–198] as illustrated in Table 4. Statistically signi
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C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 0.2 LET

Mρ=2.5 TeV

1.6 TeV 1.2 TeV

Im (FT)

0.1

0.0

-0.1

95% C.L.

-0.2 0.9

1

1.1

1.2 Re (FT)

1.3

1.4

Fig. 14. Sensitivity of a 500 GeV LC with 500 fb−1 of data to M T via the W -boson form factor; from Ref. [194]. The predicted values for a 2.5-TeV T or non-resonant scattering (LET) now lie outside 95% c.l. curve for the prediction of the Standard Model. Table 4 Total numbers of W + W − → 4-jet and ZZ → 4-jet signal S and background B events calculated for a 1:5 TeV LC with integrated luminosity 200 fb−1 Channels

SM mH = 1 TeV

Scalar MS = 1 TeV

Vector MV = 1 TeV

LET

S(e+ e− → ((W U +W −) B(backgrounds) √ S= B S(e+ e− → ((ZZ) U B(backgrounds) √ S= B S(e− e− → ((W − W − ) B(backgrounds) √ S= B

160 170 12 120 63 15 27 230 1.8

160 170 12 130 63 17 35 230 2.3

46 4.5 22 36 63 4.5 36 230 2.4

31 170 2.4 45 63 5.7 42 230 2.8

Results are shown for the SM, for strongly-coupled models with scalar or vector resonances and for low-energy-theorem + − (LET) scattering. Events are summed over the mass range 0:5 ¡ MWW ¡ 1:5 TeV except √ for the W W channel with a narrow vector resonance in which 0:9 ¡ MWW ¡ 1:1 TeV. The statistical signi
whereas the backgrounds increase by smaller factors. A 2 TeV e+ e− linear collider would increase the signal rates by roughly a factor of 2–2.5. When both the pair production and gauge boson fusion processes are taken into account, the direct signal of a strongly-coupled symmetry-breaking sector is generally stronger at a high-energy linear collider than at the LHC; Fig. 15.

C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402

Signal Significance

Mρ = 1240 GeV

Signal Significance

Mρ= 1600 GeV

10 2

10 2

10

10

1

10 2

LHC

LC

500

LC

1000

LC

1

1500

M = 2500 GeV ρ

LHC

LC

500

LC

1000

LC

1500

LET 10 2

10

1

279

10

LHC

LC

LC

LC

500

1000

1500

1

LHC

LC

LC

LC

500

1000

1500

Fig. 15. Direct strong symmetry breaking signal signi
A strongly-coupled electroweak symmetry breaking sector can also be investigated [134] at photon– photon colliders in the modes → ZZ, → W + W − , → W + W − W + W − , and → W + W − ZZ. While irreducible backgrounds in the
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If the TC group is SU (NT ), the anomalous coupling between the PNGB and the gauge bosons G1 and G2 is given, as for the QCD pion, by [99,208,209] NT AG1 G2

g1 g2 H%(1 k1% k2( H11 H2 ; 22 FT

(2.76)

where NT is the number of technicolors, AG1 G2 is the anomaly factor (see below), the gi are the gauge couplings of the gauge bosons, and the ki and Hi are the four-momenta and polarizations of the gauge bosons. The model-dependent value of the anomaly factor is calculated in [99,208,209]. The dominant production mode for a PNGB at LEP I was generally the Z → P a process [204] which has a branching ratio of order 10−5 EZ 0 → Pa = 2:3 × 10

−5

123GeV GeV FT

2

NT A Z 0

2

MZ20 − MP2a MZ20

3 :

(2.77)

The
MQ ¡ 2mt :

MP ¿ 2mt ; (2.78)

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281

Fig. 16. Projected 95% c.l. exclusion reach for electrically neutral PNGB as a function of mass and number of technicolors at a 1 TeV LC [214].

Fig. 17. Projected 5-sigma discovery reach for electrically neutral PNGB as a function of mass and number of Technicolors at a 1 TeV LC [214].

was studied in Ref. [214]. The signal events will stand out from the background due to the hard monochromatic photon recoiling against the PNGB decay products. Based on a Monte Carlo simulation of signal and background, Ref. [214]
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PNGB decays to the heaviest available fermion pair, the production and decay chain will look like [214] U + W − g; e+ e− → PT0 8 g → t tUg → bbW U → bbg;

MP8 ¡ 2mt :

MP8 ¿ 2mt (2.79)

Characteristics that will help distinguish the signal events from those of the qqg U background are monochromaticity of the gluons in the signal events are monochromatic and their large spatial separation from the PNGB decay products. The left-right asymmetries and t tU spin correlations will also diQer from those of the background distributions. Based on a Monte Carlo simulation of signal and background, Ref. [214] asserts that the electrically neutral color-octet PNGB can be excluded or discovered for any realistic values of NTC provided that MP ¡ 2mt . For higher-mass PNGB,
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283

Fig. 18. Projected P 0 production rate (solid), rate including PNGB decay to bbU (dashed), and irreducible → bbU √ background (dotted) for collisions at s = 500 GeV with integrated luminosity of 20 fb−1 [170].

3. Extended technicolor 3.1. The general structure of ETC A realistic Technicolor model must address the Iavor problem. That is, it must incorporate a mechanism for generating quark and lepton masses and the various weak mixing angles, together with CP-violation. This implies that the ordinary quarks and leptons of the Standard Model need to couple to the technifermion condensate that breaks the electroweak symmetry. There must also be a mechanism for violating the conserved techni-baryon quantum number: techniquarks must be able to decay, since stable techni-baryonic states are cosmologically problematic [64]. The classic way to ful
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therefore required, coupling to the new ETC gauge bosons. In a full theory we assume a large U QU and U . A simple gauge group GETC which contains all of the desired currents of the form QQ, generalization is a blockwise imbedding of SU (NTC ) into SU (NETC ) with NETC ¿ NTC . Clearly NETC must be large enough to accommodate representations containing both ’s and Q’s. For example, consider how this imbedding might work in the minimal model. The minimal model includes one weak doublet of technifermions (ND = 1) and a typical choice of Technicolor group NTC = 4. ETC must couple the technifermions to all 12 left-handed electroweak doublets of ordinary quarks and leptons and all 12 singlets. Then we should think of the ordinary fermions of each generation as falling into 3 doublets of quarks and one of leptons under the gauged SU (2)L and an analogous set of doublets under a global SU (2)R . This yields an ETC scheme with a gauge group GETC = SU (16) × SU (2)L × SU (2)R × U (1). Both SU (NTC ) and SU (3)c are imbedded into SU (16); hypercharge arises as Y=2 = I3R + (B − L)=2, where B − L is a diagonal generator in SU (16). The fermions form two fundamental 16 multiplets; one of them: (Qc ; Qk ; Qm ; Qy ;

1 r;

1 g;

1 b;

2 r;

c; k; m; y = Technicolors;

2 g;

2 b;

3 r;

3 g;

r; g; b = colors

3 b;

1 lep ;

2 lep ;

3 lep )L

(3.80)

is a doublet under the electroweak group SU (2)L , and a singlet under SU (2)R , a (2; 1); the other is a (1; 2). The Technicolor condensate will therefore break both the SU (2)L and the SU (2)R . For the Farhi–Susskind model, an analogous ETC extension might be to a gigantic gauge group placing all of the quarks and leptons and techniquarks (including each Technicolor copy) into a single multiplet of a compact gauge group, e.g., SU (56). Here the full Standard Model gauge interactions themselves would be subgroups of the master group SU (56). Clearly a large variety of models are possible, and we will describe several general approaches to building consistent models in this Section (several variant schemes are also reviewed in [57]). Starting from a high-energy theory based on a master gauge group GETC , it is necessary to arrive at a low-energy theory in which the only surviving gauge groups are those of Technicolor and the Standard Model. The group GETC has generators, T a , which form a Lie Group, [T a ; T b ] = ifabc T c :

(3.81)

The Technicolor gauge group GTC must be a subgroup of GETC , since the ’s do not carry Technicolor, while the Q’s do. Then GETC must undergo symmetry breaking at a scale ETC down to its subgroup GTC : GETC → GTC × · · ·

at ETC ;

(3.82)

where the ellipsis denotes other factor groups, including perhaps the full Standard Model SU (3) × SU (2) × U (1). This leaves the Technicolor gauge bosons (generators denoted T˜ a ) massless, and elevates all of the coset ETC gauge bosons to masses of order ETC . Indeed, the breaking may proceed in a more complicated way, e.g., it may occur in n steps GETC → G1 → G2 → · · · Gn−1 → GTC . A class of models in which this symmetry-breaking sequence is achieved dynamically is known as “Tumbling Gauge Theories” [219–225]. Typically, at each step, the subgroup will evolve and become strong, permitting new condensates to form, which further break the theory into the next

C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402

Q

285

Q

q

q

Fig. 19. The exchange of an ETC gauge boson allows a standard quark (or lepton) to communicate with the techniquark U condensate, QQ.

subgroup. This occurs repetitively at diQerent scales, and may in principle be a way of generating the mass hierarchy for quarks and leptons of diQerent Iavors. Another possibility is that there are fundamental scalars associated with supersymmetry at some high energy scale, and the breaking of GETC (whether single or sequential) can be driven by a Higgs mechanism (see Section 3.7). 3.1.2. Low energy relic interactions While the only elements of the original ETC gauge group that survive at low energies are the generators of the Technicolor and Standard Model gauge groups, the low-energy phenomenology of these models includes additional eQects caused by the broken ETC generators (Fig. 19). On energy scales % . ETC exchange of the heavy ETC bosons corresponding to those broken generators produces three types of eQective contact interactions among the ordinary and technifermions:

Uab

U % TU a U % TU b U % TU a QQ U % TU b Q U U a U % TU b Q Q U ab Q % T + + U ab 2ETC 2ETC 2ETC

:

(3.83)

Here the , and are coeVcients that are contracted with generator indices and their structure depends upon the details of the parent ETC theory. In the TU a we include chiral factors such as (1 ± 5 )=2 (i.e., the theory is Iavor-chiral and the ETC generators have diQerent actions on left- and right-handed fermions). We can now Fierz rearrange these operators to bring them into the form of products of scalar and pseudo-scalar densities. Upon Fierz rearrangement, we can pick out the generic terms of greatest phenomenological relevance:

ab

U a QQT U bQ QT QU L T a QR U R T b + ab 2ETC 2ETC

L

+ ab

U LT a

U

R RT 2ETC

b

L

+ ··· :

(3.84)

Note that after Fierzing we must include the identity matrix among the generators, which we do by extending the range of the generator indices to include zero: 1 ≡ T0 . As a consequence of these generic terms we see that the physical eQects of ETC go beyond generating the quark and lepton masses and mixings. On the positive side, ETC interactions, such as the -terms, can elevate the masses of some of the light Nambu–Goldstone bosons, such as the techniaxions, to
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π

π

Fig. 20. ETC gauge boson exchange across a fermion loop. This is an term insertion leading to PNGB mass.

On the negative side, Extended Technicolor produces the -terms, four-fermion contact interactions amongst ordinary fermions of the same Standard Model gauge charges. This leads generally to Iavor-changing neutral current eQects in low energy hadronic systems, e.g., it can lead to dangerously large contributions to the KL KS mass diQerence. Lepton number violation will also generally occur, leading to enhanced rates for % → e + and related processes. It is a diVcult model-building challenge to limit these dangerous eQects while generating adequately large quark and lepton masses to accommodate the heavier fermions, e.g., charm, =, b, and especially t. It is, moreover, important that the oblique electroweak corrections involving TC and ETC eQects be under control [226], as will be discussed in Section 3.2. These and other 19 phenomenological considerations have had a large inIuence on ETC model-building. We will look at explicit examples of models constructed to address such questions in Section 3.3. 3.1.3. The -terms: techniaxion masses The four-technifermion terms (with coeVcients ) can potentially solve a problem we encountered in the previous discussion of the minimal and Farhi–Susskind models in Section 2. Loops involving

term insertions, which represent ETC gauge boson exchange across a technifermion loop, with external PNGBs (Fig. 20) generally induce masses for the PNGBs. This mechanism can elevate the masses of the undesirably light PNGBs to larger values more consistent with experiment. To see this explicitly, we can make use of an NJL or Georgi–Manohar model of the techniquark condensate, which instructs us to replace the techniquark bilinears in Eq. (3.84) with the corresponding chiral <elds (a discussion of the chiral dynamics of ETC may be found in [168,231]). Speci
˙

˙

Sab ≡ exp(ic T˜ c =FT )ba ;

(3.85)

where dotted indices are for SU (2)L and undotted are for SU (2)R and the T˜ a are generators of the Technicolor subgroup of GETC . This leads, through the diagram of Fig. 20, to a PNGB interaction term in the eQective Lagrangian: ∼ 19

2 6TC

ab c2 NTC Tr(ST a S† T b ) : 2ETC

(3.86)

There has recently been a great deal of interest in the possibility of a small departure of the observed g − 2 of the muon from the Standard Model expectation [227,228]. Some authors claim that it is diVcult for TC (or Topcolor) to yield comparable eQects [229], while others claim that it is not [230].

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Expanding Eq. (3.86) in the a T˜a we see that the induced technipion mass terms are ∼−

2 6TC

ab c2 NTC Tr([c T˜ c ; T a ][T b ; d T˜ d ] : 2ETC FT2

(3.87)

Those technipions associated with the Technicolor generators, T˜ a that commute with ETC, i.e., for which [T˜a ; T b ] = 0 ;

(3.88)

will have vanishing mass contributions from ETC. Technipions associated with the noncommuting generators, on the other hand, will generally receive nonzero ETC contributions to their masses of order ∼ NTC 2TC =ETC . This situation is akin to that of QCD: the T a are analogues of the electric charge operator, and a ˜ T are the generators of nuclear isospin. The neutral pion of QCD is associated with a generator, I3 that commutes with the electric charge, Q = I3 + Y=2, hence the neutral pion receives no contribution to its mass from electromagnetism. The charged pions, associated with generators I1 ± iI2 , which do not commute with Q, receive nonzero electromagnetic contributions to their masses [167]. Among the PNGBs that arise in the Minimal and the Farhi–Susskind models, the techniaxions can receive masses, at best, only from ETC. In classic ETC the prospects for mass generation have been treated in detail in [99]; a typical result is 1 2 Maxion ≈ (a few GeV)2 : (3.89) NTC We will see subsequently that electrically charged PNGBs receive masses from ETC that are small compared to their electromagnetic contributions. Similarly the colored states, P3 and P8 of the Farhi–Susskind model receive larger masses from QCD than from conventional ETC. We will discuss below, however, an additional dynamical ingredient of modern ETC theories known as “Walking”. In Walking Technicolor, any eQective operator involving the techniquarks can be signi
NTC 3TC : 2ETC

(3.90)

This would seem to liberally allow a
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-b

L

T ETC Z

T b

L

Fig. 21. Direct correction to the ZbbU vertex from exchange of the ETC gauge boson that gives rise to the top quark mass. Technifermions are denoted by ‘T’.

Higher-order eQects of the terms also yield observable consequences. The key example for ETC model building is Rb . In a classic ETC model, mt is generated by the exchange of an electroweak-singlet ETC gauge boson of mass METC ∼ gETC ETC coupling with strength gETC . At energies below METC , ETC gauge boson exchange may be approximated by local four-fermion operators, and mt arises from an operator coupling the left- and right-handed currents: −

2 gETC ( U iL % TLiw )(UU wR % tR ) + h:c : 2 METC

(3.91)

where T = (U; D) are technifermions, and i and k are weak and Technicolor indices. Assuming there is only a single weak doublet of technifermions: mt =

2 3 gETC U U ≈ TC : U 2 METC 2ETC

(3.92)

The ETC boson which produces mt through Eq. (3.91) also necessarily induces the related operator [232]: −

2 gETC ( U iL % TLiw )(TU iL % 2 METC

iw L )

+ h:c :

(3.93)

This aQects the Zbb vertex as shown in Fig. 21, which indicates the ETC boson being exchanged between the left-handed fermion currents (with T ≡ DL since the ETC boson is a weak singlet). This process alters the Z-boson’s tree-level coupling to left-handed bottom quarks gL = (− 12 + 1 sin2 ,W )(e=sin ,W cos ,W ) by [232] 3 -gLETC = −

e mt e )2 TC 1 (I3 ) = )2 ; 2 2 ETC sin , cos , 4 TC sin , cos ,

(3.94)

where the right-most expression follows from applying Eq. (3.92). Here ) is a mixing parameter angle between the W and Z and the ETC bosons of the theory. Such shifts in gL alter the ratio of Z boson decay widths Rb ≡

U ^(Z → bb) : ^(Z → hadrons)

(3.95)

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This is a convenient observable to work with since oblique [233,235,114–116] and QCD corrections largely cancel in the ratio, making it sensitive to the direct vertex correction. One
(3.97)

are generally induced, and these give new contributions to experimentally well-constrained quantities [240]. For example, the
fK2 m2K . 10−14 ; 2ETC

(3.98)

where we might expect ∼ sin2 ,c ∼ 10−2 in any realistic model. Hence, we obtain ETC & 103 TeV :

(3.99)

The second term of Eq. (3.97) induces the lepton-Iavor-changing process % → eee; U ; e ; from this, we estimate a somewhat weaker bound, ETC & 101 TeV [173]. Applying the bound Eq. (3.98) to our expression for ETC-generated fermion masses (3.90), and assuming ∼ ∼ , yields an upper bound on the masses of ordinary quarks and leptons that

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a generic ETC model can produce (we use TC . 1 TeV, . 10, NTC . 10): mq; ‘ . NTC

3TC . 100 MeV : 2ETC

(3.100)

Hence, producing the mass of the charm quark is already problematic for a classic ETC model. A remedy for production of the charm and, marginally, the b-quark masses is “Walking Technicolor”, as we discuss below (Section 3.4). We remark that another way to suppress the ab eQects is to construct theories based upon SO(N ) groups containing ETC and Technicolor, with fermions in the real N representation. These can avoid Iavor-changing neutral current-like interactions in the tree approximation [241,242], a consequence of the representations that can be generated by products of the fundamental N representation at that order. An operator such as Eq. (3.97) will be generated in loops, and one still obtains a limit of order 1 GeV on the allowed fermion masses. Ultimately addressing the heavy top quark requires new dynamical mechanisms such as noncommuting ETC (Section 3.3.2) or “Topcolor” (Section 4.2). These -term problems, it should be noted, have analogues in the Minimal Supersymmetric Standard Model, under the rubric of “The SUSY Flavor Problem”. The mass matrices of the squarks and sleptons may in general require diagonalization by diQerent Iavor rotations than those of the quarks and leptons. This leads to similar unwanted FCNC and lepton number violating processes. The typical constraints upon mass diQerences between
T

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291

100

m

h mt 1000

S Fig. 22. Fit of the precision electroweak data to the Standard Model plus the S, T parameters described in the text (from Peskin and Wells [243]). The ellipse is the
to be mt = 174:3 GeV and mH = 100 GeV. In Fig. 22, we present the 68% con<dence contour (1:51) for a current S–T
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ZT ∼ −0:3. The Farhi–Susskind model has a full generation of 3 colors each of techniquarks and a “fourth color” of leptons, and thus yields ZS ∼ 0:8 with ZT ∼ −0:3, and would appear to be ruled out. In the minimal model we might be able to pull ourselves from the point (0:2; −0:3) back into the experimentally favored ellipse by adding a positive contribution to ZT ∼ 0:4 from additional new physics. To return to the ellipse in a Farhi–Susskind–like model would require large negative sources of S as well. There are possible physical sources of the additional contributions to S and T which are required to make Technicolor consistent with the measured oblique parameters. First, strong dynamics models that address the Iavor problem must usually include new higher-dimension operators from physics at higher scales that will aQect the oblique parameters; a general analysis has been made by a number of authors [114,247–249]. The possible existence of such operators obviates any claim that new dynamical models are ruled out by the S–T constraints. Second, the presence of precocious (∼ TeV scale) extra dimensions can play an important role, as recently discussed by Hall and Kolda, [250]. Third, Appelquist and Sannino have shown that the ordering pattern for vector-axial hadronic states in SU (N ) vector-like gauge theories close to a conformal transition need not be the same as predicted in QCD-like theories. As a consequence, the S-parameter in near-critical technicolor theories can be greatly reduced relative to QCD-like theories [251]. Fourth, speci
M2 N log 2 ; 3 m

where 1 N =− 12

(j+ + 1) (j− + 1)

(3.102) 2

− 1 j− (j− + 1)(2j− + 1) :

(3.103)

If the particles with the smallest j− are the lightest, then the multiplet can yield negative contributions to S. Large values of |ZS| can be obtained from multiplets with large weak isospin. Several avenues for producing small or even negative contributions to S have been discussed in the literature. Models with Majorana neutrino condensates naturally produce negative contributions to S [254], which can in principle compensate large positive contributions. Luty and Sundrum [255] constructed models in which the PNGBs give negative contributions to S. To obtain ZS −0:1 from this source, one needs technifermions with jL = 2 and PNGBs with masses of order 200 GeV. 20

The custodial SU (2) is essentially the global SU (2)R acting on right-handed fermions, and broken by U (1)Y and the Higgs–Yukawa couplings. It insures that when the electroweak symmetries are broken, there remains an approximate SU (2) global symmetry. This global SU (2) symmetry implies, in a chiral Lagrangian, that f+ = f0 .

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293

More generally, other models may include arbitrary numbers of vectorlike pairs of electroweak charged particles [256] (e.g., both a left-handed and right-handed gauged doublet with j− = 0 and a Dirac mass coming from other external dynamics) with no cost in S. The Top–Seesaw model (Section 4.4) contains additional vectorlike fermions, allowing both S and T to be completely consistent with present data. Likewise, the Kaluzsa–Klein recurrences of ordinary quarks and leptons in extra-dimensional models are vectorlike pairs, and are largely unconstrained by S. These examples are discussed in Section 4.6. It is straightforward to engineer large positive contributions to ZT [257]. Particles with masses much larger than ∼ 1 TeV can contribute to ZT if their masses have an up-down Iavor asymmetry. The contribution is of order 2 mU − m2D : |ZT | ∼ 2 (3.104) mU + m2D |mU − mD | can typically be at of order 100 GeV. One class of theories that generate this kind of contribution to ZT are the Technicolor models in which ETC gauge interactions are replaced by exchange of weak-doublet techni-singlet scalars [258–260] (see Section 3.5). The eQects of new Z bosons on oblique corrections were studied by many authors (see [243] and references therein). Such particles occur in new strong dynamical models, e.g. as in Topcolor Models. Peskin and Wells investigate the region of parameters for any Z 0 model in which the shifts due to the Z 0 compensate those of a heavy Higgs boson. Models with extra dimensions can also exhibit “compensation” between eQects of new vector bosons and heavy Higgs bosons. Rizzo and Wells [261] have shown that the 95% C.L. bound on the Higgs boson mass could reach ∼ 300 to ∼ 500 GeV for Kaluzsa–Klein masses MKK in the range ∼ 3 to ∼ 5 TeV. The coupling of the new sector is typically strong (see Section 4). In this vein, Chankowski et al. [262] has argued that two-Higgs-doublet models (which can be viewed as eQective Lagrangians for composite Higgs models) can be made consistent with the electroweak
This could be engineered, for instance, by including many vectorial techniquarks which would aQect the function without adding to S.

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To summarize, the potential conIicts between models of new strong dynamics and the tight experimental constraints on the size of the oblique electroweak parameters can be averted in several ways. Whether it is particle content, the details of Technicolor dynamics, or the behavior of higher-dimension operators from higher-energy physics which comes to the rescue, it is clear that strong dynamics models are not a priori ruled out. The challenge for model-builders is to produce concrete models that are consistent with the constraints. 3.3. Some explicit ETC models Having introduced the broad theoretical outlines and main experimental challenges of ETC, we now present some explicit models. Our summary emphasizes strategies employed in model-building and does not represent a complete list of all models. This discussion will anticipate some of the ideas of Walking ETC which will be more fully treated in Section 3.4. 3.3.1. Techni-GIM An essential problem in ETC is to
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295

+ 0 NLC e e -> γ P

LHC

pp -> TT

LHC pp -> θθ 4 TeV Tevatron pp -> TT LEP2 e+e- -> TT Tevatron pp -> TT + 0 LEP e e -> γ P

fS-1 [GeV]

100

500

Fig. 23. The potential for probing the scale of TC-GIM interactions at present and future colliders (the high-scale models). The most important reaction(s) for each collider is(are) indicated inside the bars (from Skiba [174]; T = P3 , , = P8 in Skiba’s notation).

numbers of any generation. This means, for example, that existing limits for leptoquarks decaying only to
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TC × SU (2)W × U (1)Y ⇒ (v0 ) ; TC × U (1)EM ; where f, u, and v0 =246 GeV are the VEVs of the order parameters for the three diQerent symmetry breakings. Note that, since we are interested in the physics associated with top-quark mass generation, only tL , bL and tR must transform non-trivially under ETC. However, to ensure anomaly cancellation we take both (t; b)L and ((= ; =) to be doublets under SU (2)heavy and singlets under SU (2)light , while all other left-handed ordinary fermions have the opposite SU (2) assignment. The two simplest possibilities for the SU (2)heavy × SU (2)light transformation properties of the order parameters that mix and break the extended electroweak gauge groups are ’ ∼ (2; 1)1=2 ;

∼ (2; 2)0 ;

“heavy case” ;

(3.105)

’ ∼ (1; 2)1=2 ;

∼ (2; 2)0 ;

“light case” :

(3.106)

The order parameter ’ breaks SU (2)L while mixes SU (2)heavy with SU (2)light . In the “heavy” case [271], the technifermion condensation that provides mass for the third generation of quarks and leptons is also responsible for the bulk of electroweak symmetry breaking (as measured by the contribution to the W and Z masses). The “light” case is just the opposite: the physics that provides mass for the third generation does not provide the bulk of electroweak symmetry breaking. While this light case is counter-intuitive (after all, the third generation is the heaviest!), it may in fact provide a resolution to the issue of how large isospin breaking can exist in the fermion mass spectrum (and, hence, the technifermion spectrum) without leaking into the W and Z masses. This is essentially what happens in multiscale models [272,273] and in Topcolor Assisted Technicolor [274]. Such hierarchies of technifermion masses are also useful for reducing the predicted value of S in Technicolor models [211]. Below the scale of ETC breaking, ETC boson exchange yields four-fermion operators among third-generation quarks and technifermions [271]: 1 1 1 U % % U U U L4f ∼ − 2 ) L UL + t R TR )U L % L + T R % t R ; (3.107) ) ) fETC where ) is a model-dependent coeVcient. When the techniquark condensate forms in the LR crossterms 22 in these operators we obtain the top quark mass of order Eq. (3.94) [172,173]. In the heavy case the technifermions responsible for giving rise to the third-generation masses also provide the bulk of the W and Z masses, and we expect FT ≈ 125 GeV (which, for mt ≈ 175 GeV, implies fETC ≈ 375 GeV) [277]. Even in the light case there must also be some SU (2)heavy breaking VEV, in order to give the top quark a mass. The spectrum of non-commuting ETC models includes an extra set of W and Z bosons which aQect weak-interaction physics at accessible energies. Mixing between the two sets of weak gauge bosons alters the Zff couplings. In addition, the one-loop diagram involving exchange of the 22

The LR-interactions become enhanced in strong-ETC models, in which the ETC coupling is
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297

top-mass-generating boson shifts the coupling of bL to the Z as in Eq. (3.94); here, the techniquarks in the loop are up-type, reversing the sign of the
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uni<ed models in which some of the key theoretical and phenomenological challenges of ETC are also addressed. Giudice and Raby [241,242] have considered a uni<ed model that starts with an ETC group of SO(4) with a full standard model family of fermions in the 4 vector representation. This sector of the model is, in spirit, similar to the Farhi–Susskind model with SO(4) replacing SU (NTC ) and containing the ETC gauge interactions. At the ETC breaking scale 1ETC ∼ 103 TeV the group SO(4) breaks to SO(3) which is the true gauge Technicolor group. As mentioned above, theories based upon SO(N ) groups containing ETC and Technicolor, with fermions in the N representation, can avoid Iavor-changing neutral current-like interactions ( terms) in the tree approximation. This leads to gauge bosons with masses of order 5 TeV for third generation, and ∼ 103 TeV for
where m is the operator’s anomalous dimension.

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299

If TC is QCD-like, then the TC coupling constant (%) is asymptotically free, and falls logarithmically as (%) ˙ 1=ln(%) above the scale TC . With the anomalous dimension m ˙ (%) we see that the radiative correction is proportional to exp[ m ln(ln(%)] ∼ (ln(ETC =TC )) m . Hence the radiative corrections are power-logarithmic factors, similar to the behavior of QCD radiative corrections to the nonleptonic weak interactions in the Standard Model. If, however, (%) is approximately constant, i.e., if the TC theory exists approximately at a “conformal <xed point”, (%) = ? = 0, where ( ? ) = 0, then the radiative correction is converted ? into a power law, proportional to exp[ m ( ? )ln(%)] ∼ (ETC =TC ) m ( ) , which is a substantially larger renormalization eQect. Such a theory is not QCD-like, but is an a priori possible behavior of a Yang–Mills gauge theory. This behavior can provide signi
c ). We recapitulate the argument below. The Euclideanized Schwinger–Dyson equation for the self-energy of a fermion in Landau gauge is given by [296]

2

S(p ) = 3C2 (R)

d 4 k ((k − p)2 ) S(k 2 ) : (2)4 (k − p)2 Z(k 2 )k 2 + S2 (k 2 )

(3.109)

Typically we approximate Z(k 2 ) ≈ 1, and linearize the equation by neglecting the S2 (k 2 ) denominator term. We assume (%) ≈ c is slowly varying. Two solutions are then found: 2

S(p ) = S(%)

%2 p2

b± ;

b± = 12 (1 ± (1 − (%)= c )1=2 ) ;

(3.110)

where the critical coupling constant is c = =3C2 (R), and C2 (R) is the quadratic Casimir of the complex technifermion representation R (recall C2 =(N 2 −1)=2N for the fundamental representation).

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U operator is The normal perturbative anomalous dimension of the QQ m = 1 − (1 − (%)= c )1=2 ∼

3C2 (R) (%) : 2

(3.111)

Hence, the solution with b− corresponds to the running of a normal mass term of nondynamical origin. The solution with b+ ∼ 1 corresponds to the high momentum tail of a dynamically generated mass having the softer ∼ 1=p2 behavior at high energies. 23 Note that, at the critical coupling

(%) = c , the two solutions coincide, which is believed to be a generic phenomenon, [73,298,299]. Moreover, if we suppose that (%) = ∗ ¿ c for 0 6 % 6 ∗ , then we
∗ ∗ −1 : (3.112) S(0) ∼ exp −

c In the energy range TC 6 % 6 ETC , the large value of (%) ≈ ∗ corresponds to an anomalous dimension of order 1, making the radiative correction factor for the technifermion bilinear (3.108) of order ETC =TC . What are the implications of Walking Extended TC for the ordinary fermion masses? In classic ETC, we have seen that fermion masses typically scale as 3TC =2ETC . Since TC ≈ 1 TeV, the phenomenological constraint on ETC & 100 TeV implies mq; ‘ . 100 MeV. Walking ETC brings a large renormalization enhancement of the techniquark bilinear by a factor of order ∼ ETC =TC , so that we now have [292] mq; ‘ ∼ 2TC =ETC ∼ 1 GeV ;

(3.113)

which is large enough to accommodate the strange and charm quarks, and the = lepton. This is born out by more detailed studies which include the full ETC boson exchange in the gap equations [291–295,300,301]. On the other hand, this is barely large enough to accommodate the bottom quark and, certainly not the top quark masses. 24 Consider that, if TC has QCD-like dynamics, the value of ETC required to
Jackiw and Johnson [297] showed long ago that this solution also forms the Nambu–Goldstone pole, con
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301

What is the origin of the small function and what other eQects may arise as a consequence? Consider the one-loop function of an SU (N )TC gauge theory with Nf techniquarks in the fundamental NTC representation: 3 11 gTC 8 NTC − Nf + · · · : TC = − (3.115) 162 3 3 2 Clearly, for TC (%) = gTC =4 to walk requires having many technifermions active between the scales TC and ETC . These need not all be electroweak doublets, e.g., they may be singlets or vectorlike doublets with respect to SU (2)L , which can help suppress contributions to S. Higher tensor representations of the SU (NTC ) gauge group are also possible [166]. In general, after the ETC breaking, the fermions in the lower energy theory fall into subsets carrying (i) only (TC); (ii) (TC) × (color); (iii) (TC) × (Iavor), and so on, including both the fundamental and higher representations of TC. The technifermions in diQerent representations may condense at diQerent scales, as will be discussed in the next section. Whether walking is caused by the presence of many technifermions in the fundamental TC representation or technifermions in higher TC representations, the chiral symmetry-breaking sector is enlarged relative to that of minimal TC models. As a result, one expects a proliferation of technipions and small technipion decay constants FT v0 . At
(3.116)

and the PNGBs are now safely elevated out of harm’s way from current experiments, but left potentially accessible to the Tevatron and LHC. More generally, Walking TC is actually an illustration of the physics of chiral dynamics in the large NIavor limit. Recent interest in the phase structure of chiral gauge theories has been inspired in part by duality arguments in SUSY theories where there exist exact results for the phase structure

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of an SU (N ) gauge theory with Nf Iavors (see the review [308]). The presence of infrared <xed points of the gauge coupling appears to be fairly generic in theories with a large number of Iavors. The infrared <xed point of the strong SU (N )TC gauge theory can also arise from the interplay of the
See [311] for a discussion of the interplay of chiral breaking and con
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303

for the bulk of the EWSB [247,315]. A scheme of this type called “Topcolor Assisted TC”, or TC2, and the phenomenological implications of the new strong top dynamics will be discussed in Section 4.2. In general, the spectroscopy of a low-scale TC model can accommodate everything from a minimal model through the Farhi–Susskind structure. The low-scale spectrum will, at the very least, include light PNGBs (technipions) and techni-vector mesons. For example, with FT ∼ 60 GeV one expects MPT ∼ 100 GeV and M T ∼ 200 GeV [52]. The technipions will be resonantly produced via techni- vector meson dominance (VMD) 26 with large rates at the Tevatron, LHC, and a linear collider [316]. The technivector mesons are expected to be, in analogy to the minimal model, an isotriplet, color-singlet T , and the isoscalar partner !T . Isospin is likely to be a good approximate symmetry, so T and !T should be approximately degenerate in mass, as is the I = 1 multiplet of technipions. The enhancement of technipion masses due to walking suggests that the decay channels T → PT PT and !T → PT PT PT are probably closed. Thus, the decay modes T → WL PT and ZL PT , where WL , ZL are longitudinal weak bosons, and T ; !T → PT may dominate. Because technipion couplings to fermions, like those of scalars, are proportional to mass, one expects the most important decay modes to be PT0 → bbU ; PT+ → cbU

or

cs; U =+ (= :

(3.117)

Heavy-quark jet tagging is, then, important in searches for low-scale TC. Eichten et al. [312] have performed an extensive analysis of the phenomenological signatures ± 0 of Low-Scale TC at the Tevatron. They present simulations of pp U → ± T → WL PT and !T → −1 0 PT for the Tevatron collider with an integrated luminosity of 1 fb . For M T 200 GeV and MPT 100 GeV, the cross sections at the Tevatron are expected to be of order a few picobarns. The narrowness of the T and !T suggests that appropriate cuts, e.g., on the
qqU → + Z 0 + 0T → WL+ WL− ;

WL± PT0 ; PT± ZL0 ; WL± PT∓ ;

PT± PT0 ;

PT+ PT− :

(3.118)

For M T ∼ 200 GeV ¿ 2MPT and MPT ∼ 100 GeV, the dominant processes have cross sections of 1–10 pb at the Tevatron and ∼ 10–100 pb at the LHC. The modes with the best signal-to-background are T → WL PT or ZL PT and T !T → PT . 26

Per the discussion of Section 2.3.3, there is an induced coupling of the T to any electromagnetic current, so this also applies at e+ e− and %+ %− colliders [157]). See [159,158] about the caveats that apply when estimating production of color-octet T by vector meson dominance.

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Fig. 24. Predicted invariant mass distributions at Tevatron Run II for T signal (black) and Wjj background (grey); vertical scale is events per bin in 1 fb−1 of integrated luminosity from Ref. [312]. Dijet mass distributions (a) with kinematic selections only, (b) with the addition of topological selections, and (c) with the addition of single b-tagging; (d) W +dijet invariant mass distribution for the same sample as (c).

Fig. 24(d) shows invariant mass distributions for the Wjj system after kinematic and topological cuts and b-tagging have been imposed as discussed in Ref. [312]. A clear peak is visible just below the mass of the T , and the peaks in the dijet mass and the Wjj mass are correlated. If the T and PT exist in the mass range favored by the low-scale TC models they can be easily found in Run II of the Tevatron. The !T is likewise produced in hadron collisions via vector-meson-dominance coupling through and Z 0 . We expect !T → PT0 , Z 0 PT0 will dominate !T → PT PT . Fig. 25(a) shows the invariant mass distribution of the two highest-ET jets for signal (black) and background (grey) events, that pass certain kinematic criteria, for an integrated luminosity of 1 fb−1 . The eQect of topological cuts is seen in Fig. 25(b). Tagging one b-jet signi
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Fig. 25. Predicted invariant mass distributions at Tevatron Run II for !T signal (black) and jj background (grey); vertical scale is events per bin in 1 fb−1 of integrated luminosity. Dijet mass distributions (a) with kinematic selections only, (b) with the addition of topological selections, and (c) with the addition of single b-tagging; (d) +dijet invariant mass distribution for the same sample as (c).

a few picobarns. In the next section, we will examine existing experimental constraints on these models. 3.6. Direct experimental limits and constraints on TC Experiments performed in the past few years have had the large data samples, high energies, and heavy-Iavor-tagging capabilities required to begin direct searches for the new phenomena predicted by models of new strong dynamics. We have just examined some of the theoretical implications for preferred channels of low-scale TC models. We now summarize the status of experimental limits on the accessible scalar mesons, vector mesons, and gauge bosons, that are generically predicted by these theories. Our focus here is on the direct searches performed largely by the LEP and Tevatron experiments, and we present many of their original exclusion plots. Tables summarizing many of these results may be found in Refs. [317,194]. As appropriate, we will also comment on more indirect searches for new physics via measurements of precision electroweak observables and on the prospects for the LHC and future lepton colliders. For further discussion of technicolor searches at LHC and even higher-energy hadron colliders, see Ref. [194].

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400 500 2 [GeV/c ]

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Fig. 26. Invariant mass of the dijet system and of the W +2 jet system (with a leptonically decaying W ) in the CDF search for T → W ± PT [319]. The points are data; the solid histogram is background; the dashed histogram shows background plus the signal from a walking TC model with M T = 180 GeV and MPT = 90 GeV. The topological cuts leading to the lower
3.6.1. Searches for low-scale color-singlet techni- ’s and techni-!’s (and associated technipions) In light of the low-scale TC model, and the generic phenomenology of TC, it is useful to examine the present-day constraints that exist, mostly from LEP and the Tevatron, on the direct observables, i.e., the masses and production signatures of technivector mesons and techni-’s. Note that a number of the searches have taken advantage of the fact that the !T of low-scale models can be visible in collider experiments [272,316,312,52]. Enhancement of technipion masses by WTC can quench the decay !T → PT PT PT , resulting in the dominant mode !T → PT [272]. Techni-! decays to qq, U ‘+ ‘− and ((U can also be signi
T ; !T decaying to e+ e− [320]. The LEP collaborations L3 [321], DELPHI [322], and OPAL [323] have released preliminary results on multi-channel searches for light T and !T . U ‘(cbU in 109 pb−1 of CDF performed a counting experiment looking for T → WPT → ‘(bb; Run I data [318,319]. They selected candidate lepton plus two-jet events with at least one jet b-tagged. The presence of peaks in the Mb; jet and MW; b; jet distributions would signal the presence of the PT and T , respectively, as indicated in Fig. 26. No deviation from Standard Model expectations was observed. CDF therefore set upper limits on the techni- production cross-section for speci
C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 115

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Fig. 27. Excluded region for the CDF search for color singlet techni- search in the mode T → W ± PT [319]. PYTHIA v6.139 : MV=MA=200 GeV 150 ρT± → W± + πT0 and ±

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Fig. 28. Predicted reach of CDF in Run II for T → WT assuming MV = MA = 200 GeV [52] (note the notational variance, PT = T ).

pair of techni- and technipion masses as indicated in Fig. 27. In Run II, with a larger data sample and a doubled signal eVciency, CDF expects to explore a signi
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B(ρT,ωT→ ee) x Cross Section (pb)

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Theory (Eichten, Lane, Womersley) Mρ,ω-Mπ = 100 GeV MT = 100 GeV MT = 200 GeV MT = 300 GeV MT = 400 GeV

1

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-2

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Fig. 29. Excluded regions for the D0 search for T ; !T → e+ e− [324].

CDF also searched [318,319] for technipions in the shape of the two b-jet mass distribution in ‘ + 2-jet and 4-jet events, using 91 pb−1 of Run I data. The former topologies arise as described above; the latter can result from either T → WPT or T → PT PT where the technipions decay to heavy Iavors and the W decays hadronically. CDF notes that the upper limits (∼ 100 pb) this search sets on production of ∼ 200 GeV techni- decaying to ∼ 100 GeV technipions provides no immediate improvement in the constraints on TC models. Run II should provide signi
T → WL WL ; WL PT ; PT PT ; PT( )

(3.119)

for the following range of techni- and technipion masses 50 GeV ¡ MPT ¡ 150 GeV;

150 GeV ¡ M T ¡ 250 GeV :

(3.120)

In the WW decay channel, all decay modes of the W bosons are included. The result is that an upper limit of 0.47 pb was set at 95% c.l. on the possible increase of the e+ e− → WW cross-section due to contributions from TC.

C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402

309

Fig. 30. Projected reach of the D0 detector in Tevatron Run IIa for T ; !T → e+ e− [52].

In other decay channels, the technipions decay predominantly to bbU or bcU (the calculated branching ratios ranged from 50% to 90%) no statistically signi
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L3

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Fig. 31. The M T –MPT region excluded by L3 at 95% c.l. [321].

DELPHI e e →πTπT;πTWL e+e-→ρT(γ) : ρT→hadrons ρT→WL+WL+ -

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Fig. 32. M T − MPT region excluded by DELPHI at 95% c.l. [325]

of resonance production. Thus, CDF obtains an upper limit on (cross-section) × (branching ratio), which it compared to the predictions of the TC models of Ref. [312]. The range of excluded masses of PT and !T are shown in Fig. 34 (the exclusion region boundary is ragged because of statistical Iuctuations in the data). CDF noted that, if the channels !T → PT PT PT or !T → ZPT were open,

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excluded 250 300 350 400 450 500 550 600

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U search and Fig. 33. The 95% c.l. excluded region in the (M T ;!T ) plane from the combination of the PT+ PT− → bqUbq 0 U search. The dashed lines show the corresponding median expected exclusions for the background only the PT → bb hypothesis [323].

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Fig. 34. The 90%, 95% and 99% c.l. exclusion regions for the CDF search for ! → PT [326]. The inset shows the limit on · B for MPT = 120 GeV; the circles indicate the limit and the solid line shows the prediction from [312].

the excluded region would be reduced as shown in the
T → PT and T ; !T → T , their excluded region would increase. Future experiments should provide de
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Fig. 35. Simulated event and background rates in the ATLAS detector for T → W ± Z → ‘± (‘ ‘+ ‘− for various M T and MPT in low-scale TC models [212]; from Ref. [327].

the ATLAS collaboration [327] show that the relatively light techni–vector resonances of low-scale TC [212] are well within reach at LHC (Fig. 35). Experiments at an e collider have potential to discover and study an !T with a mass up to about ∼ 1 TeV in the processes e− → e− !T → e− Z; e− W + W − Z [328]. As shown in Fig. 36, at the stage where only detector acceptance cuts have been applied, the cross-section for an !T decaying to WWZ can be of order ∼ 10–100 fb, and as much as an order of magnitude above background (!T → Z is below background). Applying various kinematic cuts, the !T could be visible for a √ range of masses and decay widths. For example, with 200 fb−1 collected at s = 1:5 TeV, an !T of width 100 GeV can be detected at the 3 level up to a mass of about ∼ 1:3 TeV.

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313

(a)

(b)

Fig. 36. Cross-section vs e+ e− CM energy for e− → e− !T with !T → WWZ and Z and the SM backgrounds thereto. Solid (long-dashed) lines are for M!T = 0:8 (1.0) TeV. In each case, the upper curve is for WWZ and the lower is for Z . The dash-dot (dotted) line is for the SM e− WWZ (e− Z ) background. In (a), 0:25EWWZ = EZ = 5 GeV (solid) or 20 GeV (long-dashed); in (b), 0:5EWWZ = EZ = 15 GeV (solid) or 40 GeV (long-dashed) (From [328]).

3.6.2. Separate Searches for color-singlet P 0 ; P 0 As discussed in Sections 3.4 and 4, many TC models require a large number ND of weak doublets of technifermions. For a given TC gauge group SU (NTC ), the number of doublets required to make the gauge coupling gTC “walk” is ND ≈ 10, as in the models of Refs. [315,330,331]. Topcolor-Assisted TC models (see Section 4) also tend to require many doublets of technifermions [212,315]. One phenomenologically interesting consequence of the large number of doublets is the √ presence of PNGB states with small technipion decay constants: FT ≈ v0 = ND . Such states will be lighter, and have generically longer lifetimes. As outlined in Section 2.3.4, data from LEP can place limits on single production of light neutral PNGBs from a variety of TC models. One benchmark example is the Lane’s low-scale TC “Straw Man Model” (TCSM) [212,213], in which the lightest technifermion doublet, composed of technileptons TU and TD can be considered in isolation. The result is two, nearly degenerate neutral mass eigenstates, whose generators are given by PT0 ∼ TU U 5 TU − TU D 5 TD and PT0 ∼ TU U 5 TU + TU D 5 TD . As shown in Ref. [329] LEP searches for hadronically-decaying scalars produced in association with

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Table 5 Limits (from Ref. [329]) on the number of technicolors, NTC , and weak doublets of technifermions, ND , for hadronically decaying PNGBs in TCSM [212,213] models as a function of the upper bound on the PNGB mass MP0 6 T

30 GeV 60 GeV 80 GeV 100 GeV 120 GeV 140 GeV 160 GeV

√ NTC ND 6

PT0 → gg

U PT0 → bb

28(b) 67(b) 283(b) — — — —

24(b) 70(b) 25(a) 40(a) 42(a) 49(a) 68(a)

The superscripted labels indicate the data used to calculate the limits: (a) means A Pa ; (b) means A ZPa .

√ a Z or may be used to place an upper bound on the product NTC ND A (where A is the relevant anomaly factor). Inserting the value of A appropriate to a particular PNGB yields an upper √ on NTC ND . Table 5 gives these bounds as a function of PNGB mass for the cases where either U decays dominate. 2-gluon or bb √ Consider the case where MPT0 6 30 GeV and b-quark decays dominate; the limit NTC ND 6 24 applies. As a result, for NTC = (4; 6; 8; 10; 12) the largest number of electroweak doublets of technifermions allowed by the LEP data is, respectively, ND =(36; 16; 9; 5; 4). The results are very similar if the two-gluon decays of the PNGB dominate instead. How do these results accord with the requirements of Walking TC? Requiring the one-loop TC beta function to satisfy TC ≈ 0, implies that 11NTC =4 weak doublets of technifermions are needed, according to Eq. (3.115). The analysis of Ref. [329] thus shows that WTC and a very light PT0 can coexist only in models with NTC = 4; 6. Similarly, the size of the TC group is restricted to NTC 6 6 [12] if the PNGB mass is 80 GeV [160 GeV]. The results are similar if the 2-loop -function is used. As a second example, we mention the results of Ref. [329] for the WTC model of Lane and Ramana [156] whose LEP-II and NLC phenomenology was studied by Lubicz and Santorelli [207]. The model has ND =9: one color-triplet of techniquarks (NQ =1) and six color-singlets of technileptons (NL =6). Of the several neutral PNGBs in this model, the one whose relatively large anomaly factors and small decay constant (FT ≈ 40 GeV) makes it easiest to produce is PL3 ∼ NU ‘ 5 N‘ − EU ‘ 5 E‘ where the subscript implies a sum over technilepton doublets. This PNGB is expected to have a mass in the range ∼ 100–350 GeV [207]. Depending on the value of the ETC coupling between the U Ref. [329] PNGB and fermions, the dominant decay of this PNGB may be into a photon pair or bb.
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Fig. 37. LEP lower bounds on MH ± as a function of B(H + → =(). From [317].

(i) LEP limits Color-singlet electrically-charged technipions PT± with the quantum numbers of the charged scalars in two-higgs-doublet models are directly constrained by the limits on pair-production of H ± derived from LEP data. When tan is large, the charged scalars of two-higgs-doublet models decay mostly to =(; if tan is small, light charged scalars decay to cs, U but for MH ± heavier than about 130 GeV, the channel H + → t bU → WbbU dominates [332,333]. The LEP searches assume that H + → =+ (= ; csU (as consistent with the mass range they can probe), and derive limits on the rate of H + H − as a function of the branching ratio B(H + → =+ (= ). As shown in Fig. 37, the lower limit is at least 77 GeV for any value of the branching ratio. If the mass of the charged scalar is less than mt − mb , then the decay t → H + b can compete with the standard top decay mode t → Wb. Since the tbH ± coupling can be parameterized in terms of tan as [334] gtbH + ˙ mt cot (1 + 5 ) + mb tan (1 − 5 ) ;

(3.121)

we see that the additional decay mode for the top is signi
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200 180 160 140 120 100 80 60 10

-1

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Fig. 38. Limits on charged scalar mass as a function of tan . The 95% exclusion bounds from CDF and D0 studies of top decays are strong functions of tan . LEP limits from Fig. 37 are also shown. From [317].

Fig. 39. Projected Run II reach of D0 charged scalar search in t → H ± b assuming (t tU) = 7 pb.

√

s = 2 TeV,

L dt = 2 fb−1 , and

of (3:28 ± 0:33) × 10−4 [339,340], as would be true in a model with an extended Higgs sector. To the extent that this is the only new physics contribution to b → s , it implies an upper bound of order 300 GeV on the mass of the charged scalar. However, the contributions of other new particles or non-standard gauge couplings can also aQect the branching ratio, making the exact limit quite model-dependent. Weaker, and also model-dependent, bounds can also be derived from measurements of b → s and b → =(= X and from =-lepton decays at LEP [341–345].

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317

Fig. 40. CDF’s 95% c.l. exclusion region for 8 → PU 3 P3 → cc(= (= [347] for several limiting values of ZM =M (P8 )−M (P3 ) which aQect the V8 decay partial width.

3.6.4. Searches for low-scale color-octet techni- ’s (and associated leptoquark technipions) In models of Walking TC, in which enhancement of the technipion masses prevents the decay V8 → PU T PT (V8 is the color-octet techni- ), decay to dijets can dominate [291,316]. The CDF Collaboration has used 87 pb−1 of Run I data to search for the eQects of V8 ’s on the dijet and b-tagged dijet invariant mass spectra [346]. No deviations from the Standard Model backgrounds were observed. A narrow V8 with mass 350 ¡ MV8 ¡ 440 GeV is excluded at 95% c.l. using the b-tagged distribution, as may be seen in Fig. 57. The mass range 260 ¡ MV8 ¡ 470 GeV is ruled out using the untagged sample. In searches for leptoquark techni-’s P3 , CDF sets NTC = 4 and allows a relevant parameter, ZM , to take on the expected value of 50 GeV and the limiting values of 0 and ∞. More precisely, ZM is the mass diQerence between P8 and P3 , and enters the calculation of the partial width for V8 → PU 3 P3 [156]. CDF reports joint limits on the masses of the P3 and V8 . U − . The observed yield was In the
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Fig. 41. CDF’s 95% c.l. exclusion region for V8 → P3 P3 → bb(= (= [347] for several values of ZM = M (P8 ) − M (P3 ).

about ∼ 450, (500; 650) GeV assuming ZM = 0, (50; ∞); heavier P3 would decay to t(= . The lower bound from the continuum search in this channel is MP3 ¿ 122 GeV at 95% c.l. For P3 decaying to b(U= , CDF’s continuum search set the 95% lower bound MP3 ¿ 149 GeV and the technipion search excludes MP3 up to the kinematic limit (MV8 =2) for techni-V8 masses up to about ∼ 600 (650, 700) GeV assuming ZM = 0 (50, ∞), as shown in Fig. 41. Leptoquarks typically decay into a quark and a lepton of the same generation. In the TC-GIM models, the P3 ’s carry lepton and quark numbers of any generation. A leptoquark decaying into an e and d–quark has the signature of a “
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Fig. 42. Invariant mass distribution for the ppU → P80 → g signal (dashed) and background (solid) in multi-scale TC with MP = 250 GeV and F = 40 GeV [185].

of about ∼ 100, a smaller background makes the signal potentially visible, as illustrated in Fig. 42. By employing a PYTHIA-level simulation the authors of Ref. [185] were able to identify cuts on the transverse momentum and invariant mass of the photon + jet system that signi
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production MZ =

em =4 sin2 , ¿

365 GeV

ALEPH

355 GeV

OPAL

and from bb production 523 GeV MZ = em =4 sin2 , ¿ 325 GeV

ALEPH OPAL

:

(3.122)

:

(3.123)

While these are weaker than the current limits on non-commuting ETC [278] or topIavor [350,353] Z bosons from precision electroweak data (Section 3.3.2), they are complementary in the following sense. The limits from a
Exchange of the ETC boson that generates mt does not modify the Wtb vertex, because the boson does not couple to all of the required fermions: (tR ; bR ; UR ; DR ).

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Required Luminosity (fb−1)

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Fig. 43. Luminosity required to discover SU(2) Z bosons preferentially coupled to the third generation [329]. Dashed curves are 3 discovery curves for a <xed mass, while solid curves are 5 discovery curves. From bottom to top, 3 curves are displayed for Z masses of 550 GeV, 600 GeV, 650 GeV, and 700 GeV. From bottom to top, 5 curves are displayed for Z masses of 550 GeV, 600 GeV, and 650 GeV. The horizontal lines indicate luminosity targets for Run II.

third-generation leptons were large, production of unlike-sign %= pairs might be seen. Because the VH VL VL triple gauge boson coupling vanishes to leading order, di-boson production will not reveal the presence of a W or Z . 3.7. Supersymmetric and bosonic technicolor 3.7.1. Supersymmetry and technicolor Shortly after the introduction of Technicolor, Witten described the general aspects of a Supersymmetric Technicolor theory [65], while Dine et al. [66,67] constructed explicit models. Several interesting possibilities arise when the ideas of Supersymmetry (SUSY) and TC are combined. Notably, it becomes possible to raise the scale of ETC [358], and thus suppress the dangerous terms. One might even be able to achieve dynamical electroweak symmetry breaking in concert with Supersymmetry breaking. These models are unfashionable at present, relative to the MSSM which guarantees a low-mass Higgs boson accessible to Run IIb or the LHC. Nonetheless, the general idea has some theoretical merit and the models have interesting and accessible phenomenology (including the potential for light fundamental scalars). In ordinary QCD we know that a chiral condensate, qq U = 0 is dynamically generated when the gauge coupling becomes strong. In Supersymmetric QCD a quark q possesses a superpartner, denoted q, ˜ the squark. The supercharge, Q, generates a transformation on these <elds of the form: {Q; q˜† q + q† q} ˜ = qq U :

(3.124)

Hence, the existence of the fermionic condensate qq U = 0 implies that: Q|0 = 0 :

(3.125)

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Whenever Q does not annihilate the vacuum it means that SUSY is broken. The <eld q˜† q + q† q˜ then becomes a massless Goldstone fermion, a Goldstino. Thus, SUSY breaking and dynamical EWSB can have an intimate connection through fermion condensation. In addition to the key papers mentioned above, we refer the interested reader to further works on Supersymmetric Technicolor, Refs. [359–362,67], and to Ref. [363], which discusses the combination of Topcolor with SUSY. 3.7.2. Scalars and technicolor: bosonic technicolor The key advantage of introducing fundamental scalars into TC is to provide an alternative to ETC. Most of the constraints and problems of ETC can be dismissed by assuming general masses and couplings of the fundamental scalar sector. We thus consider theories in which the additional <elds that communicate the TC condensates to the ordinary quarks and leptons are fundamental scalars. We view these scalars as ultimately associated with Supersymmetry, where their masses can be protected by fermionic chiral symmetries [364–369,360]. Alternatively, these scalars may, in principle, be bound states arising within a high energy strongly coupled theory [275,276]. Conceivably, they can also be viewed as relics from extra dimensions, such as Wilson lines (Section 4.6). In what follows, we momentarily disregard the details of the higher-energy physics that enables the scalars to have masses of order the weak scale and consider the key features and phenomenology of TC models with fundamental or eQective scalars. We classify the models according to the weak and TC charges of the scalar states. (i) Weak-doublet techni-singlet scalars We begin by discussing TC models whose spectrum includes one or more weak-doublet TC-singlet scalars. These correspond to a natural low-energy limit of strongly-coupled ETC models [275]. Moreover, the presence of a weak-doublet techni-singlet scalar has a suVciently large eQect on the vacuum alignment of the technifermion condensate to make an SU (2) TC gauge group viable [370]. In the minimal model of this type [258], one adds to the Standard Model gauge and fermion sectors a simple SU (N ) TC sector, with two techniIavors that transform as a left-handed doublet WL = (pL ; mL ) and two right-handed singlets, pR and mR , under SU (2)W , with weak hypercharge assignments Y (WL ) = 0, Y (pR ) = 1=2, and Y (mR ) = −1=2. The technifermions and ordinary fermions each couple to a weak scalar doublet ' = ('+ '0 )T which has the quantum numbers of the Higgs doublet of the Standard Model. Unlike the Standard Model Higgs doublet, ' has a nontachyonic mass (M'2 ¿ 0) and is not the primary source of electroweak symmetry breaking. The scalar has Yukawa couplings to the technifermions: ˜ + pR + WU L '1− mR + h:c: L'T = WU L '1

(3.126)

When the technifermions condense, these couplings cause ' to acquire an eQective VEV, ' = f ≈ 41T f3 =M'2 . Because the scalar couples to ordinary fermions as well as technifermions, the ordinary fermions obtain masses from the diagram sketched in Fig. 44 mf ≈ 1f h

4f3 ; M'2

(3.127)

where h is the scalar coupling to the ordinary fermions. The coupling matrices 1f are proportional to the mass matrices mf and drive Iavor symmetry breaking. The quark Iavor symmetries are broken

C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 -

TL

φ

T

R

f f

323

R

L

Fig. 44. Fermion–technifermion interaction through scalar exchange. This diagram is responsible for fermion mass generation once the technifermions condense.

according to the same pattern as in the Standard Model so that the quarks mix via the usual CKM matrix and the standard GIM mechanism prevails. Both the technipion decay constant and the scalar VEV contribute to the electroweak scale: f2 + 2 f = v02 . The technipions and the isotriplet components of ' mix. One linear combination becomes the longitudinal component of the W and Z; the orthogonal one remains in the low-energy theory as an isotriplet of physical scalars [371]. The coupling of the charged physical scalars to quarks has the same form as in a type-I two-Higgs doublet model [371]. Imposing the requirement that the isoscalar component of ' have no VEV enables one to eliminate f and f and calculate observables in terms of M' , h and 1. This model has been studied in the limits (i) in which 1 is negligible and one works in terms of (M' ; h) [258,259,371], and (ii) in which M' is negligible and one works in terms of (1; h) [260,371]. The leading Coleman–Weinberg corrections to the scalar potential have been included in studies of both limits. The phenomenology of TC models with weak-doublet scalars has been found to be in agreement with experiment. Models of this kind do not produce unacceptably large contributions to K 0 –K 0 or B0 –B0 mixing, nor to the electroweak S and T parameters [258–260]. In addition, the new scalars in the model can be made heavy enough to evade detection, even in the limit where the scalar doublet is assumed to have a vanishing SU (2) × U (1) invariant mass [260]. There are negative corrections to Rb , but they never exceed −1% in the regions of parameter space allowed by other constraints; likewise, the rate for b → s is less than in the Standard Model, but not so altered as to conIict with experiment [371]. Current bounds on the parameter space of the model [372] are shown in Figs. 45 and 46; in creating the plots, the parameters M' (limit i) and 1' (limit ii) have been eliminated in favor of mass of the physical isoscalar state, m . Supersymmetrized models with weak-doublet techni-singlet scalars were introduced in Refs. [364,365] and aspects of Iavor physics and renormalization group evolution were studied in Refs. [366–369]. The minimal such model [364,365] contains all the <elds of the Minimal Supersymmetric Standard Model (MSSM), a set of SU (N )TC gauge bosons and their superpartners, and color-singlet technifermion super<elds transforming under SU (N )TC × SU (2)W × U (1)Y as TUR ≡ (NTC ; 1; 1=2);

TDR ≡ (NTC ; 1; −1=2);

TL ≡ (NTC ; 1; 0) :

(3.128)

The superpotential includes WSSM from the MSSM and an additional part WHTC from Yukawa couplings of the two Higgs super<elds HU and HD to the technifermion super<elds T : WHTC = gU HU TUR TL + gD HD TDR TL : R-parity is imposed, just as in the MSSM; the technifermions transform like matter.

(3.129)

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Fig. 45. Constraints on technicolor with scalars in limit (i), where the scalar self-coupling is negligible, plotted in the physical basis (m , h). The allowed region of parameter space (shaded) is bounded by the contours m = 114 GeV (solid), Rb − RSM b = 0:26% (dashes) and hf = 4f (dot-dash). Other contours of constant Rb are shown for reference. The current bound from searches for charged scalars mp± = 79 GeV is shown (long dashes) along with the reference curve m ± = mt − mb . The constraint from B0 BU 0 mixing is labeled “B-line” [372]. p

The Higgs <elds have positive squared masses in the perturbative vacuum. The superpotential terms WHTC , however, produce terms linear in the Higgs <elds in the Lagrangian when the technifermions condense [364,365]. This causes the Higgs <elds to acquire VEVs: HU = gU TUR TUL =m2HU ; 3

HD = gD TDR TDL =m2HD :

(3.130)

Thus for TT ∼ (600 GeV) , gU ∼ 1 and mHU ∼ 1 TeV, one obtains HU ∼ 100 GeV, yielding a realistic top quark mass. If mHD ¿ mHU , the VEV of HD will be much smaller than that of HU , producing the required top-bottom mass splitting. In variant models with larger technifermion content, WHTC generates technifermion current-algebra masses, yielding masses of order 200 GeV– 1 TeV [364,365] for the technipions in the spectrum. These supersymmetrized TC models minimize the FCNC problems that usually aQect SUSY and TC models [366–369]. In the minimal version, scalar exchange among quarks and leptons generates no tree-level FCNC. If additional Higgs multiplets are introduced to explain the hierarchy of fermion masses and mixings, tree-level FCNC may be re-introduced; scalar masses of order ∼ 10 TeV suVce to suppress them. Furthermore, because the lightest Higgs bare mass mHU is of order ∼ 1 TeV, superpartner masses of order ∼ 1–10 TeV become natural, reducing the degree of low-energy squark and slepton mass degeneracy required to avoid undue FCNC from loops. (ii) Weak-singlet technicolored scalars Theories which include weak-singlet technicolored scalars address the intergenerational fermion mass hierarchy more directly. In the early model of Ref. [373], exchange of technicolored

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325

Fig. 46. Constraints on technicolor with scalars in limit (ii), where the scalar mass is negligible, plotted in the physical basis (m , h). Curves labeled as in Fig. 45. From [372].

weak-singlet scalars induces four-fermion interactions between a trio of technifermions and one fermion. Ordinary fermions then mix with technibaryons and become massive. Unfortunately, this model predicts unacceptably large FCNC and tree-level contributions to the T parameter. A more natural account of the mass hierarchy is provided by models [369,374,360,375] in which exchange of technicolored weak-singlet scalars induces four-fermion interactions between pairs of fermions and technifermions. The small CKM elements associated with the third-generation quarks arise because the tL and bL mass eigenstates are automatically aligned. However, because the scalar can couple to both tR and bR , the top-bottom mass ratio must be provided by a ratio of Yukawa couplings. We will discuss three models with interesting features in more detail. A supersymmetric model with technicolored weak-singlet scalars and some interesting mass phenomenology was introduced by Kagan [369]. The gauge group is SU (NTC ) × SU (3) × SU (2) × U (1); three families of quark and lepton super<elds and two sets (T1 ; T2 ) of the color-singlet technifermion super<elds as in Eq. (3.128) are present. The Higgs sector, consists of a vector-like pair of color U 1; −1=6) triplet Higgs <elds transforming under SU (NTC ) × SU (3) × SU (2) × U (1) as !c ≡ (NTC ; 3; c and ! ≡ (NTC ; 3; 1; 1=6) and a vector-like pair of color singlet <elds ) ≡ (NTC ; 1; 1; 1=2) and ) ≡ (NTC ; 1; 1; −1=2). While only two generations of quarks receive masses at tree-level in this model, gaugino and sfermion exchange enable the
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TU

i L

TU x

Q

hi

R

~ g

i

ωc

UL

ω

hi

UR

(a)

~ g

ωc x ω

U

x

x

UL

UL

x

UR

UR

(b)

Fig. 47. (a) Tree-level mass for two generations of fermions due to technifermion condensation. (b) Radiative mass contribution from squark and gluino exchange. Analogous contributions to charged lepton masses arise from exchange of ) and electroweak gauginos.

techniscalar-generated chromomagnetic moments for the quarks and several scenarios for quark mass generation in non-supersymmetric versions of the model are discussed in [369,376]. Dobrescu has studied a related supersymmetric TC model with only one doublet of technifermions [360]. Only the third generation fermions acquire large masses and the other two generations’ smaller masses are generated radiatively by sfermion and gaugino exchange. The predicted rates of neutral K and B meson mixing, CP violation in KL and B meson decays, and CP asymmetries in B decays or ZS = 1 transitions, are all consistent with experiment. Dobrescu and Terning [375] have also shown that a TC model with technicolored weak-singlet scalars can produce a negative contribution to the S parameter. This model includes the Standard Model <elds, an SU (NTC ) technicolor gauge group, two Iavors of technifermions 3R ≡ (PR ; NR ) ≡ (NTC ; 1; 2; 0), PL ≡ (NTC ; 1; 1; 1), and NL ≡ (NTC ; 1; 1; −1) and three technicolored scalars transformU 1; −1=3), !t ≡ (NTC ; 3; U 1; −7=3), and !b ≡ (NTC ; 3; U 1; 5=3). The couplings of the ing as ' ≡ (NTC ; 3; t Z boson to the (t; b)L doublet (-gL ) and the tR and bR quarks (-gR ; -gRb ) are found to be shifted by amounts depending on the Yukawa couplings and the scalar masses. The model therefore predicts that the value of the S parameter is reduced from the typical one-doublet technifermion contribution by an amount depending on those couplings: S ≈ 0:1NTC − 1:02(3-gL − 2-gRt + -gRb ) :

(3.131)

(iii) Weak-doublet technicolored scalars Models incorporating weak-doublet technicolored scalars [370] can explain not only the intergenerational fermion mass heirarchy, but also some elements of the intragenerational mass hierarchies. These models [370] include the Standard Model gauge and fermion sectors together with a minimal TC sector. This takes the form of an asymptotically free SU (NTC ) gauge group, which becomes strong at a scale of order 1 TeV, and one doublet of technifermions which transform under the SU (NTC )×SU (3)C ×SU (2)W ×U (1)Y gauge group as 3L ≡ (PL ; NL )=(NTC ; 1; 2)0 , PR =(NTC ; 1; 1)+1 and NR = (NTC ; 1; 1)−1 . The large top quark mass is generated by a scalar multiplet, &t , which transforms under the SU (N )TC × SU (3)C × SU (2)W × U (1)Y gauge group as: (NTC ; 3; 2)4=3 . Its Yukawa interactions may be written without loss of generality as Lt = Cq qL3 NR &t + Ct 3L tR i2 &t † + h:c: ;

(3.132)

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where qL3 ≡ (tL ; bL ) is the left-handed weak eigenstate t −b quark doublet, and the Yukawa coupling constants, Cq and Ct , are de
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mass and electroweak scale through the Pagels–Stokar formula. It was subsequently elaborated by Marciano [387,388], and Bardeen et al. [383]. The latter paper provides a detailed analysis of the scheme with the connection to, and improvement by, the full renormalization group. The Higgs boson is shown to be a “deeply bound state” [389] composed of tUt, where the top and Higgs masses are predicted by the quasi-infrared <xed point [379,380]. The NJL model for top quark condensation must be considered as an approximation to some supposed new strong dynamics. Indeed, the discussion of models only becomes complete when a concrete proposal for the new dynamics is given, e.g., a new gauge interaction, “Topcolor”, [304] which we discuss in the next section. Technically, the Standard Model can always be rewritten as an NJL model for any fermion, by a suVciently arbitrary and wide range of choices of higher dimension operators at the composite scale, as emphasized by Hasenfratz et al. [390]. For example, combining certain d = 8 operators with the d = 6 NJL interaction, restricting oneself to the fermion bubble approximation, and choosing coeVcients for these operators to be absurdly large, of order ∼ 106 , one can argue that the Higgs is composed of the electron and positron! In doing this, however, one is perversely tuning cancellations of the coupling of composite Higgs boson to its constituents in the infrared, a phenomenon one does not expect in any natural or realistic dynamics. In Topcolor the additional operators are under control, and one can estimate the coeVcients to be small, - O(1). Indeed, Topcolor was invented [304] to address the criticism raised in Ref. [390]. There are proposals other than Topcolor, based upon strongly coupled U (1) or non-gauge interactions, which also aim to provide a concrete basis for the new strong interaction, see Refs. [391–393]. As these interactions are not asymptotically free, it is diVcult to understand how they become strong in the infrared or uni<ed at high energies. Note that U (1) interactions, moreover, are typically subleading in Nc and the NJL model is a large-Nc approximation. Topcolor thus has an advantage; we will have a new dynamics such as Topcolor in mind throughout the following discussion. We can implement the notion of top quark condensation by adapting the Nambu–JonaLasinio (NJL) model following Refs. [385,386]. A new fundamental interaction associated with a high energy scale, M , involving principally the top quark, is postulated as a four-fermion interaction potential: V ∼−

g2 U a ( tRa )i (tURb M2 L

b i

) + ··· ;

(4.133)

where (a; b) are color indices and (i) is an SU (2)L index. This is viewed as a cut-oQ theory at the scale M , (i.e., the interaction presumably softens due to topgluon exchange above this scale). For any g2 this interaction is attractive and will form a bound-state boson, H ∼ U L tR . With suVciently large (supercritical) g2 ¿ gc2 the interaction will trigger the formation of a low energy condensate, H ∼ tUt . The condensate has the requisite I = 1=2 and Y = −1 quantum numbers of the Standard Model Higgs boson condensate, enabling it to break SU (2) × U (1) → U (1) in the usual way (Fig. 48). The subsequent analysis is a straightforward application of the NJL model, as discussed in Appendix B. For supercritical g2 ¿ gc2 the theory undergoes spontaneous symmetry breaking. The associated Nambu–Goldstone modes become the longitudinal W and Z. In the fermion–loop approximation, one obtains the Pagels–Stokar formula which connects the Nambu–Goldstone boson decay

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329

Fig. 48. Fierz rearrangement of the attractive Nambu–Jona-Lasinio interaction contains the color current–current interaction to leading order in 1=Nc .

constant, f = vwk , to the constituent quark mass, [385,386,383]: M2 Nc 2 2 2 f = vwk = m log 2 + k : 162 t mt

(4.134)

Here mt is the top quark mass, which is now the dynamical mass gap of the theory. The constant k is associated with the precise matching onto, and de
M2 2Nc − (M 2 − %2 ) : 2 g (4)2

(4.135)

Taking % → vwk , we see that, in order to have M vwk ∼ mH , the coupling constant of the NJL theory must be extremely
(4.136)

This implies extreme proximity of g2 to the critical value gc2 = 82 =Nc . If M ∼ 1015 GeV then the coupling must be
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with the “compositeness boundary condition”: Nc 1 → ln(M 2 =%2 ); 2 (4)2 gt

%→M :

(4.138)

The boundary condition states that if the Higgs doublet is a pure tUt bound state, then the Higgs– Yukawa coupling to the top-quark must have a Landau pole at the composite scale M . The low energy value of gt , as given by the solution to Eq. (4.137), is the quasi-infrared <xed point, given by the approximate vanishing of the rhs of Eq. (4.137). The <xed point is only ln(ln(M )) sensitive to the UV scale M . The low energy quasi-infrared <xed point prediction in the Standard Model of mt ≈ 220 GeV with M ∼ 1015 GeV [378,379] is large in comparison to the observed mt = 176 ± 3 GeV. The result does, however, depend on the exact structure (e.g., the particle content) of the high energy theory, and one comes suVciently close to the physical top mass that perhaps new dynamics can be introduced into the Standard Model to <x the prediction. In the MSSM, for example, one obtains mt ∼ 200 sin GeV, which determines a predicted tan when compared to the experimental mt [382,394]. Top quark condensation has been adapted to supersymmetric models [395], but a problem arises. Essentially, the nonrenormalization of the superpotential implies that the eQective bound state Higgs mass does not run quadratically; it runs only logarithmically owing to the Kahler potential, or kinetic term, renormalization eQects. Hence, the requisite new strong interaction scale M must be of order vweak , and cannot be placed in any sensible way at ∼ MGUT . This may be acceptable in the context of top-seesaw models (Section 4.4), but the general viability of Supersymmetry in these schemes has not been examined. While the quasi-infrared <xed point does play an important role in the MSSM [382,394], it appears that the naive compositeness interpretation is diVcult to maintain in SUSY. There have been many applications of these ideas to other speci
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331

4.2. Topcolor 4.2.1. Gauging top condensation Our previous discussion of top condensation noted that the relevant interaction Lagrangian, Eq. (4.133), must be viewed as an eQective description of a more fundamental theory. Let us presently consider what that theory might be. A key observation is that a Fierz rearrangement of the interaction term leads to [304]: g2 U a g2 1A i 1A b i % U − 2 ( L tRa )i (tURb ) = 2 tR + O(1=Nc ) ; tUR (4.139) iL % M M 2 L 2 where Nc = 3 is the number of colors. This is exactly the form (including the sign) induced by a massive color octet vector boson exchange, and suggests a new gauge theory with certain properties: (i) it must be spontaneously broken at a scale of order M ; (ii) it must be strongly coupled at the scale M to produce deeply bound composite Higgs bosons and trigger chiral condensates; (iii) it must involve the color degrees of freedom of the top quark, analogous to QCD. The relevant models therefore involve embedding of QCD into some large group G which is sensitive to the Iavor structure of the Standard Model [304,274]. From this point of view the embedding of QCD into a minimal SU (3)1 × SU (3)2 gauge group at higher energies seems a plausible scenario. 28 The second (weaker) SU (3)2 gauge interactions act upon the
We will see in Section 4.5 that such a structure occurs in theories with extra dimensions, and anticipates the idea of “deconstruction”.

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the spectrum as an isovector multiplet of PNGBs, ˜ a . These objects acquire mass as a consequence of the interference between the dynamical and ETC masses of the top quark, i.e., the masses of the ˜ a will be proportional to H. 29 We refer to the ˜ a as top-pions (note that in the minimal pairing of the L = (t; b)L doublet with the tR singlet there is no singlet top-< ). For H - 0:05 − 0:10, we will
Technicolor itself can be a spontaneously broken theory with NJL-like dynamics [68]. This has the bene
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333

Common solutions involve introducing additional strong U (1) interactions that are attractive in U channel [304]. This can take the form of imbedding the weak the tUt channel and repulsive in the bb hypercharge, U (1)Y → U (1)Y 1 × U (1)Y 2 , since Y has the desired properties. There are then constraints on the running of the strong U (1) couplings and the demands of tilting. EQective <eld theory analysis of tilting indicates it is not an unreasonable possibility, but may require some
(4.140)

where SU (3)1 × U (1)Y 1 (SU (3)2 × U (1)Y 2 ) couples preferentially to the third (
(t; b)R ∼ (3; 1; (4=3; −2=3); 0) ; =R ∼ (1; 1; −2; 0) ;

(u; d)L ; (c; s)L ∼ (1; 3; 0; 1=3)

(u; d)R ; (c; s)R ∼ (1; 3; 0; (4=3; −2=3)) ;

((; ‘)L ‘ = e; % ∼ (1; 1; 0; −1)

‘R ∼ (1; 1; 0; −2) :

(4.141)

The extended color interactions must be broken to the diagonal subgroup which can be identi<ed with QCD. We assume this is accomplished through an (eQective) scalar <eld: U y; −y) : . ∼ (3; 3;

(4.142)

In fact, when . develops a VEV the theory spontaneously breaks down to ordinary QCD ×U (1)Y at a scale assumed to be ∼ 1 TeV, before Topcolor becomes con
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and the U (1) <eld Z% ), are then de
1 1 1 = + ; g12 q12 q22

(4.144)

where g3 (g1 ) is the QCD (U (1)Y ) coupling constant at MTC . We demand cot ,1 and cot , 1 to tilt the strongest interactions and to select the top quark direction for condensation. The masses of the degenerate octet of topgluons and Z are given by MB ≈ g3 M=sin , cos ,MZ ≈ g1 M=sin , cos , . The usual QCD (U (1)Y electroweak) interactions are obtained for any quarks that carry either SU (3)1 or SU (3)2 triplet quantum numbers (or U (1)Yi charges). Integrating out the heavy bosons Z and B gives rise to eQective low energy four fermion interactions. The eQective Topcolor interaction mediated by B is strongest in the third generation and takes the form: A A A 2N 1A 1 1 1 % % U % U LTopC = − 2 tU % t + b b tU t + b b ; (4.145) 2 2 2 2 MB where N=

g32 cot 2 , : 4

(4.146)

U channels and invariant under color SU (3) This interaction is attractive in the color-singlet tUt and bb and SU (2)L × SU (2)R × U (1) × U (1) where SU (2)R is the custodial symmetry of the electroweak interactions. In addition we have the U (1)Y 1 interaction Lagrangian (which breaks custodial SU (2)R ): 2 2N1 1 U 2 1 U 1 U U t R % tR − bR % bR − ‘L % ‘L − =UR % =R LY 1 = − 2 ; (4.147) L % L + 3 3 2 MZ 6 where g12 cot 2 , (4.148) 4 and where L = (t; b)L , ‘L = ((= ; =)L and N1 is assumed to be O(1). Note that while too small a value for N1 signi<es
by using the large-Nc NJL (fermion loop) approximation (similar criticality constraints were discussed previously in the NJL model of top condensation of Miransky et al. [385,386]).

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335

+ <ττ>

κ1 6.0

(B)

4.0 symmetric (A) 2.0 + (C) 1.0 2.0 κ Fig. 49. The full phase diagram of the Topcolor model. The top quark alone condenses for values of N and N1 corresponding to hatched region. In other regions additional unwanted condensates turn on (for still larger N1 a =b U condensate forms. (A) MB = 1:0 TeV; (B) MB = 3:0 TeV; (C) upper bound from Z → == U ( The phase diagram of the model is shown in Fig. 49 from [407]. The criticality conditions (4.149) de 336 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 color group to its QCD subgroup are “Iavor-universal colorons” coupling to all quark Iavors with equal strength [420]: LC = g3 cot ,(C A · JCA ) ; (4.150) where JC% A = q q U % 1A q : 2 (4.151) The weak and hypercharge assignments of the quarks and leptons are as in Classic Topcolor, so that the properties and couplings of the Z boson are unchanged. Once again, the values of Ni = gi2 =4 are jointly constrained by several diQerent pieces of physics. Above all, the top quark should be the only fermion to condense and acquire a large mass. Applying the gauged NJL analysis [422–425] to all the Standard Model fermions, one can seek solutions with U = 0 for f = t. Top condensation occurs provided that 30 [421,407] tUt = 0 and ff 2 2 4 4 N3 + (4.152) N1 ¿ − s − Y : 27 3 3 9 where we again take Ncrit; NJL ≈ 2=3 [34,35]. Quarks other than the top quark will not condense provided that additional limits on N1 and N3 are satis<ed. In Classic TC2 models, the limit comes U = 0, while in Iavor-universal TC2 models, an even stronger limit comes from from requiring bb ensuring cc U = 0 [421]: N3 + 2 Y2 2 4 4 ¡ − s − Y : 27 N1 3 3 9 (4.153) As shown in Fig. 50, the phase diagram of the Iavor-universal model includes a region in which only top condensation occurs; the region is similar to that of classic TC2 (see also discussion of phase plane in [424,425]). Several types of physics further constrain the allowed region of the N1 –N3 plane. As already mentioned for Classic TC2, mixing between the Z and Z bosons alters the predicted value of the Z decay width to tau leptons from the standard model value [407,421]. The or T parameter is also sensitive both to Z–Z mixing [277] and to single coloron exchange across the top and bottom quark loops of W and Z vacuum polarization diagrams [426,421]. Finally, if the Landau pole of the strong U (1) interaction is to lie at least an order of magnitude above the symmetry-breaking scale , then N1 - 1 [421]. Fig. 50 summarizes these constraints on the N1 − N3 plane for Iavor-universal TC2 models. The constraints from Z and the Landau pole also apply to the phase diagram in Fig. 49. In the end, both kinds of Topcolor models are con Likewise, if one applies the fermion bubble approximation to leptons, The strong U (1) interaction will not cause leptons to condense if [421,407] N1 ¡ 2 − 6 Y . This approximate condition is superceded by other constraints. C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 κ1 337 < ττ>(3)=/ 0 6 expt. excludes (4) 5.0 5 κ1 4 (1) 3.0 3 =/ 0 (2) = 0 2 (5) 1.0 1 κ3 (6a) (6b) (6c) 0 1.4 1.5 1.6 1.6 1.7 κ3 1.8 1.8 1.9 2 2.1 2.0 Fig. 50. Joint constraints on coloron and Z couplings (from [421]). Curves (1) – (3) outline the ‘gap triangle’ where only tUt = 0 in Iavor-universal TC2 models; in ordinary TC2 models, the triangle is roughly symmetric about the lowest point. The region above curve (4) is excluded by data on X ∗ ; the region above curve (5) is excluded by data on Z → =+ =− . Lines (6a– 6c) are upper bounds on N1 that hold if the U(1) Landau pole lies one, two, or ε mt Fig. 51. The interference of the explicit breaking term ˙ H, and the dynamical mass ∼ mt yields an estimate of the top-pion mass. explicit mass ∼ Hmt [304,274]. The top-pion decay constant is estimated from the Pagels–Stokar formula; using M = MB ∼ 1:5 TeV and mt = 175 GeV, it is f ≈ 60 GeV. The Lagrangian for the coupling of the top-pions to quarks takes the form: i i U gbt ˜ itU 5 t ˜ 0 + √ tU(1 − 5 )b˜ + + √ b(1 (4.154) + 5 )t ˜ − 2 2 √ and the coupling strength is gbt ˜ ≈ mt = 2f [274]. This Lagrangian, written above in the current basis, will in general contain generational mixing when one passes to the mass-matrix eigenbasis. Estimating the induced top-pion mass from the fermion loop yields m2˜ = NHm2t MB2 HMB2 ; = 82 f2 log(MB =mt ) (4.155) 338 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 where the Pagels–Stokar formula is used for f2 (with k =0) in the last expression. For H=(0:03; 0:1), MB ≈ (1:5; 1:0) TeV, and mt = 175 GeV this predicts m˜ = (180; 240) GeV. We would expect that H is subject to large renormalization eQects and, even a bare value of H0 ∼ 0:005 consistent with ETC, can produce sizeable m˜ ¿ mt . Charged top-pions as light as ∼ 165 GeV, would provide a detectable decay mode for top quarks [314]. Burdman has discussed potentially dangerous eQects in Z → bbU resulting from low mass top-pions and decay constants as small as ∼ 60 GeV [239]. A comfortable phenomenological range is slightly larger than our estimates: m˜ % 300 GeV and f % 100 GeV. These values remain subject to large uncertainties. The b receives a mass of ∼ O(1) GeV from ETC. Remarkably, however, it also obtains an induced mass from instantons in SU (3)1 . The instanton eQective Lagrangian may be approximated by the ‘t Hooft Iavor determinant, 31 which is the eQective Lagrangian generated by zero-modes when instantons are integrated out. The instanton induced b-quark mass can then be estimated as 3kmt m? ∼ 6:6 kˆ GeV (4.156) b ≈ 82 where we generally expect kˆ ∼ 1 to 10−1 as in QCD, from Note that the SU (3)1 CP-angle, ,1 , cannot be eliminated from the full quark mass matrix because of the ETC contribution to the t and b masses. Indeed, it can lead to induced scalar couplings of the neutral top-pion, and an induced CKM CP-phase [407]. C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 339 the third and det .† . . + c32 TU L cR H + c31 TU L uR H 3 M0 M0 M0 +c23 CU L tR H.† det .† + c22 CU L cR H + c21 CU L uR H M04 +c13 UU L tR H.† det.† + c12 UU L cR H + c11 UU L uR H + h:c: M04 (4.158) Here T = (t; b), C = (c; s) and U = (u; d). The mass m0 is the dynamical Topcolor condensate top mass. Furthermore det . is de where we have kept only terms of O(H) or larger. This is a triangular matrix (up to the c12 term). When it is written in the form UL DUR† with UL and UR unitary and D positive diagonal, restrictions on UL and UR may be inferred. In the present case, the elements UL3; i and ULi; 3 are vanishing for i = 3, while the elements of UR are not constrained by triangularity. Analogously, in the down quark sector DLi; 3 = DL3; i = 0 for i = 3 with DR unrestricted. The situation is reversed when the opposite corner elements are small, which can be achieved by choosing H ∼ (1; 1; −1; 0; 12 ). The full CKM matrix is, as usual, given by K = UL DL† . 340 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 b d π b d Fig. 52. Top-pion contribution to B0 –BU 0 mixing. d b b d Fig. 53. Top-gluon contribution to B0 –BU 0 mixing. Triangularity, thus implies that either the matrix UL has large oQ diagonal elements, while UR has small oQ diagonal elements (we will denote this case (U : 10)), or vice versa, (U : 01). Without triangularity (U : 11) is allowed; with exact Iavor symmetry we have (U : 0). These are not C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 341 by about two orders of magnitude. This is close to the current experimental limit. Thus, D0 − DU 0 mixing emerges as an intriguing probe of topcolor. Data showing that the rate of D0 − DU 0 mixing is close to the SM prediction would pose a severe diVculty for topcolor models. There are various other eQects and limits on the models that can be obtained in mostly B and D, and to a lesser degree in K physics. Many of these are sensitive to the Z which is more model dependent. We refer the interested reader to some of the relevant literature, beginning with [407,427,428]. The measurement of Rb in Z → bbU also yields an important constraint, [429,239,430]. This is subject to larger uncertainties because subleading 1=Nc top-pion loops dominate the leading large Nc top-gluon loops. The degree of uncertainty can be seen by examining the analysis of [239] and noting that in their exclusion 342 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 Fig. 54. Top-Higgs production cross-section via gluon fusion times branching ratio to tUc at √ s = 2 TeV [431]. Fig. 55. Mass distribution for the tUc jet system for mht =200; 300 GeV and for the Wjj and Wbb backgrounds at [431]. √ s=2 TeV models (Section 4.4), r = 1 so that the CP-even scalar decays mainly to VV ; however, there can also be a light CP-odd scalar that is unable to decay to VV and therefore decays to tUc instead [246]. As shown in Fig. 54, the production cross-section times Iavor-changing branching ratio at Run IIb can provide several hundred top-higgs events even for Utc ∼ 0:02. Since the top-higgs is expected to be rather narrow, the signal peak should be visible for scalar masses up to the top threshold (see Fig. 55). At LHC, the production rate is dramatically enhanced since the primary production mode is gluon–gluon fusion; the cross-section times branching ratio is about 100 times that shown in Fig. 54 for the same values of Utc . Recent studies are optimistic for this process [433], but detailed background studies are required. Recent work [434–437] also suggests that ˆ0 ’s may be visible through their tUc decays at a LC running in e+ e− or collision mode. The cross-section for e+ e− → tUc or → tUc at a linear C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 343 (a) (b) Fig. 56. s-channel charged top-pion production with the benchmark Yukawa couplings given in the text. For reference, the dashed curves show the results for Yukawa couplings satisfying the 3-sigma Rb bound [438]. collider receives a contribution of order 20 fb from neutral top-pions; an integrated luminosity of 50 fb−1 might make the top-pion visible [435]. On the other hand, the processes e+ e− → t ; Zt , with t → tUc, have cross-sections of order 1.5 and 0:3 pb, respectively for top-pion masses below √ 350 GeV [434]. A LC with s = 1 TeV would produce of order 400 (100) tUc (Z tUc) events due to top-pions, potentially making both channels observable. (ii) Charged top-pions In the presence of sizeable Iavor-changing couplings tUc of the ˆ0t , a large Iavor-mixing coupling for the charged top-pions cb U t+ can be induced [438]. This enables charged top-pions to be singly produced at sizeable rates in the s-channel at hadron colliders [439] or photon–photon linear colliders [440]. Ref. [438] has calculated the production cross-sections for several charged top-pion masses at a variety of colliders using the following benchmark set of couplings (typical for topcolor models): √ L R UtbL = Ucb = 0, UtbR = 5Ucb , gtb = 3 2mt =v. As shown in Fig. 56, the cross-sections for Run II and the LHC are 4.0 and 0:55 pb, respectively; those at photon–photon colliders are also sizeable. Hence, the authors of [438] suggest that charged top-pions would be visible up to masses of 300 –350 GeV at Run II and 1 TeV at LHC, and that light enough top-pions would also be detectable at a linear collider running in the photon–photon mode. In contrast, the work of Ref. [216] suggests that it will be diVcult to detect the eQects of charged top-pions on single top production at an e collider in the process e → tUb(. 344 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 10 3 10 3 10 10 2 1 10 Excluded: 280 < M < 670 1 200 400 600 200 10 Γ/M=0.5 10 Excluded: 340 < M < 640 1 200 400 600 2 M(g T ) (GeV/c ) - 10 2 σ • Br {gT → bb } (pb) - σ • Br {gT → bb } (pb) 3 c) Topgluons • CDF 95% CL Limit 400 600 2 M(g T ) (GeV/c ) New Particle Mass (GeV/c2) 10 Γ/M=0.3 - Technirho / Topcolor Z / Standard Z Vector Gluinonium b) Topgluons • CDF 95% CL Limit σ • Br {gT → bb } (pb) 10 2 - σ • Br {X → bb } (pb) a) Narrow Resonances • CDF 95 % CL Limit 3 d) Topgluons • CDF 95% CL Limit’ 10 2 Γ/M=0.7 10 Excluded: 375 < M < 560 1 200 400 600 2 M(g T ) (GeV/c ) Fig. 57. CDF search for topgluons in bbU [346]. We mention that in titling scenarios and variant schemes, there will be analogous bottom-pions U For a discussion of the associated phenomenology of these objects see which couple strongly to bb. [441,442,417]. 4.3.2. Colorons: new colored gauge bosons Several of the dynamical models we have discussed include extended strong interactions SU (3)1 × SU (3)2 , with coupling constants h2 h1 . A general prediction of such models is the existence of a color-octet of massive gauge bosons (colorons). The best search strategy for colorons depends on how they couple to the diQerent quark Iavors. We will discuss the phenomenology resulting from the two most common choices for the quarks’ charge assignments. (i) Topgluons In models such as Topcolor [304], TC2 [274,407], or Top Seesaw [245], only the third-generation quarks transform principally under the stronger SU (3)2 group. As a result, the massive topgluons (B%a ) couple predominantly to third-generation quarks [304]. The topgluons are expected to be heavy (M 0.5 –2:0 TeV) and broad (E=M 0:3– 0.7) resonances. In production at, e.g., the Tevatron, in the dominant process qq U → (g; B) → tUt the amplitude involves (g3 tan ,) × (−g3 cot ,) so that the factors involving , cancel (and there is characteristic destructive interference above threshold). Thus, , aQects the rates only through the decay width of B%a . Topgluons of moderate mass may be produced directly at the Tevatron [187,443]. CDF has recently used their measured upper limit on cross-section time branching ratio of new resonances decaying to bbU to place a limit on topgluons [346] (Fig. 57). They exclude topgluons of width EB = 0:3 M in the mass range 280 ¡ M ¡ 670 GeV, of width 0:5 M in the range 340 ¡ M ¡ 640 GeV, and of width EB = 0:7 M in the range 375 ¡ M ¡ 560 GeV. A simulation of topgluon production and decay combined with an extrapolation of the CDF b-tagged dijet mass data from Run I [194] indicates C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 Events/25 GeV/c 2 92.2% 10.8% 6 4 2 200 (a) 345 600 400 2 mtt (GeV/c ) 200 (b) 600 400 2 mtt (GeV/c ) Fig. 58. Invariant mass distribution for top pairs: pairs: D0 data (histogram), simulated background (triangles), simulated S+B (dots). In (a) mt unconstrained; in (b) mt = 173 Gev [445–448]. Fig. 59. PT distribution for hadronically-decaying tops in lepton+jets events from CDF [449]. that in Run IIb, the topgluon discovery mass reach in bbU It has been noted, e.g. that a narrow 500 GeV Z boson is inconsistent with the observed shape of the high-mass end of CDFs Mtt distribution [451]. 346 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 20 cot2 θ 15 ∆ρ excludes 10 D0 excludes Mc/cot 5 > 837 GeV 0 1 CDF excludes 2 3 4 Mc (TeV/c2) Fig. 60. Experimental limits on Iavor-universal colorons from sources described in the text [455]. The horizontally hatched region at large cot , is not part of the Higgs phase of the model [420]. measured to ±11 (6; 5)%. Topgluons of mass up to 1:0–1:1 TeV (1:3–1:4 TeV) would be visible in 2 fb−1 (30 fb−1 ) at Run IIb, either in a search for a new broad resonance, or through their eQects on the magnitude of the t tU total cross-section. Projected Run IIb limits on · B for new resonances decaying to t tU are illustrated in Fig. 64. For further details on current and future topgluon searches at the Tevatron, see Table III of Ref. [194]. The situation at the LHC and VLHC has been recently studied in considerable in [443], conclude that tUt is overwhelmed by QCD backgrounds. (ii) Flavor-universal colorons In theories in which all quarks carry only SU (3)2 charge [420,331,421], the massive colorons couple with equal strength to all quark Iavors (Section 4.1.2(ii)). As a result, they will appear in dijet production and heavy Iavor production at hadron colliders. A comparison of these processes [453] has indicated that the limits on colorons from dijet production should be the more stringent ones. Existing limits on Iavor-universal colorons from several sources are summarized in Fig. 60. Exchange of light strongly-coupled colorons across quark loops would cause large contributions to the T parameter; accordingly, the region (M=cot ,) ¡ 450 GeV is excluded. 33 Light narrow colorons would have been seen by a Run I CDF search for new particles decaying to dijets (see [455]); this excludes the cross-hatched region at low cot 2 ,. The light-shaded region (M=cot , ¡ 759 GeV) is excluded by the shape of the dijet angular distribution measured by D0 [456]. Finally, the shape of the dijet mass spectrum measured by D0 [450] sets the strongest limit: M=cot , ¿ 837 GeV [455]. In the context of Iavor-universal TC2 models, the value of N3 must be approximately 2 in order for the colorons to help cause the top quark to condense as discussed earlier [421]; this is equivalent to cot , ≈ 4. Hence, in these models, the limit set by D0 on the coloron mass is Mc ¿ 3:4 TeV. Run IIb will, naturally, be sensitive to somewhat heavier colorons. According to the TeV 2000 report the limit on · B will improve as shown in Fig. 61. The predicted · B for a coloron of cot , = 1 is equivalent to that for an axigluon (shown); larger values of cot , increase the rate. 33 A C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 347 Fig. 61. Anticipated 95% c.l. upper limit on · B for new particles decaying to dijet as a function of the new particle mass for various integrated luminosities at the Tevatron. Axigluon curve corresponds to a coloron of cot , = 1 [452]. 4.3.3. New Z bosons As discussed in Section 4.1, both Classic TC2 [274,407] and Flavor-Universal TC2 [331,421] models include an extra U (1) group and predict the presence of a massive Z boson. The couplings of this Z to fermions are generally not Iavor-universal, and the more strongly suppressed the Z couplings to (4.161) where the coupling of the 348 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 Optimal Z’ 14000 MHZ_hL 12000 10000 8000 6000 4000 2000 0.2 0.4 5000 0.6 sin^2 chi 0.8 1 Optimal Z’ MHZ_hL 4000 3000 2000 1000 0.025 0.05 0.075 0.1 0.125 0.15 0.175 sin^2 chi 0.2 Fig. 62. Lower bounds at 95% c.l. on mass of “optimal” non-universal Z boson [278]. Solid curve is from a global (i) Z and precision tests Studies of precision electroweak limits on the Z bosons [278,457] in these models have obtained lower bounds on MZ as a function of sin2 &. The lower bound is found to depend quite strongly on the hypercharge assignments of the fermions because the shift in the Z boson’s coupling to fermions depends directly on Yh : H e sin , f f 2 f 1 + ; (4.163) − sin &Y Y -g ≈ − h x cos , cos2 & sin2 & where H ≡ 2f2t =v2 (YhtL − YhtR ). In models with a so-called “optimal” Z , in which the third generation fermions have Yh = Y; Y‘ = 0 and those of the other two generations have Yh = 0; Y‘ = Y , the Z can be as light as 630 GeV (see Fig. 62) at 95% c.l. for sin2 & ≈ 0:0784; a Z mass less than a TeV is allowed for 0:0744 - sin2 & - 0:0844 [278]. In the TC2 model of Lane [330], where the fermions have considerably larger Yh charges (see Table 6), the lower bound on the Z mass is correspondingly higher, about 20 TeV. The strong constraint here comes from sensitivity to atomic parity violation in Cs [330]. A variant of Lane’s model [458] in which the lepton’s U (1)h couplings are vectorial C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 349 Table 6 Fermion charges for Lane’s model [330] 1st, 2nd Yh 3rd Yh (u; d)L ; (c; s)L u R ; cR d R ; sR (ve ; e)L ; (v% ; %)L e R ; %R −10:5833 −5:78333 −6:78333 −1:54 2.26 (t; b)L tR bR (v= ; =)L =R 8.7666 11.4166 10.4166 −1:54 2.26 has a Z limit much closer to that of the “optimal” scenario (though lacking the low-mass region near sin2 & = 0:0784) [457]. The non-universal couplings of the Z boson to fermions cause it to induce lepton number violating processes. A study in Ref. [459] has shown that % − e conversion in nuclei is about an order of magnitude better than the decay % → 3e [460] for constraining the magnitudes of the lepton mixing angles. The decay % → e yields weaker bounds. Present data allows the Z mass to be as large as 1 TeV and the magnitudes of the lepton mixing angles lie roughly between the analogous CKM entries and their square-roots [459]. Looking to the future, Ref. [461] 350 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 20 CDF 95% C.L. Upper Limits for Γ = 0.012MX σX • Br{X→tt} (pb) CDF 95% C.L. Upper Limits for Γ = 0.04MX Leptophobic Topcolor Z′, Γ = 0.012MZ′ 10 9 8 7 6 Leptophobic Topcolor Z′, Γ = 0.04MZ′ 5 4 3 2 1 0.9 0.8 0.7 0.6 0.5 400 500 600 700 800 900 1000 2 MX (GeV/c ) Fig. 63. The 95% c.l. upper limits on X · BRX → t tU as a function of mass (solid and open points) compared to the cross section for a leptophobic topcolor Z (thick solid and dashed curves) for two resonance widths [451]. Min σ∗B(X→tt) for a resonance to be observed at the 5σ level. 1 fb-1 1 -1 σ∗B (pb) 10 fb -1 -1 100 fb 10 -2 10 TopColor Z‘, Γ=1.2% TopColor Z‘, Γ=10% 400 500 600 700 800 900 1000 Mtt GeV/c2 Fig. 64. Anticipated [452] Run IIb limits on · B(X → t tU) and predictions for a topcolor Z . C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 351 sensitivity to · B in the Mtt distribution, and the corresponding predictions for topcolor Z bosons. The search by the CDF Collaboration for new resonances decaying to bbU [346], which was previously mentioned (Section 4.3.2, Fig. 57) as setting limits on technirhos, falls slightly short of setting a limit on topcolor Z bosons. In addition to pursuing Z bosons in these heavy quark channels, the Run IIb will also be able to look for Z bosons decaying to = leptons as described in Section 3.6.5. In the case with the lowest standard model background, Z → == → e%((, U a Z boson with a mass up to about 600 GeV should be accessible (depending on mixing angle) [329]. For further details on current and future Topcolor Z boson searches at the Tevatron, see Table IV of Ref. [194]. A Z boson would also be visible in the process e+ e− → == at an NLC [421]. Assuming a 50% eVciency for =-pairs and requiring an excess over standard model backgrounds of identifying == == == (N − NSM ) ¿ 5 NSM , it appears that the eQects of a 2:7 TeV Z boson with 1 tan2 & 6 1 would √ be visible in a 50 fb−1 sample taken at s = 500 GeV. A 1:5 TeV NLC with 200 fb−1 of data would be sensitive to Z bosons as heavy as 6:6 TeV. The other channels suggested earlier in the context of topIavor models might also be useful for 352 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 Note that the Top Seesaw model, unlike TC2, does not invoke Technicolor; rather, it replaces Technicolor entirely with Topcolor. It oQers novel model building possibilities for attacking the Iavor problem [465–467,417]. Extensions have been constructed in which the W , Z, t and b all receive dynamical masses. Construction of an explicit model of all quark and lepton masses seems plausible [465,466]. Because the & quarks need not carry weak-isospin quantum numbers, and enter in vectorlike pairs, 34 the constraints on the number of techniquarks from the S parameter are essentially irrelevant for the Top Seesaw. The Top Seesaw also (4.167) 1 m21 = [m20 + M&2 + %2 − (m20 + M&2 + %2 )2 − 4%2 m20 ] 2 ≈ 34 (4.166) m20 %2 + O(m40 %4 =M&6 ) M&2 + m20 + %2 See Refs. [417,467] for models where the &’s are weak isodoublets. (4.168) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 and 353 1 m22 = [m20 + M&2 + %2 + (m20 + M&2 + %2 )2 − 4%2 m20 ] 2 ≈ M&2 + m20 + %2 + O(m20 %2 =M&2 ) ; (4.169) where limits for large M& are indicated. The fermionic mass matrix thus admits a conventional seesaw mechanism, yielding the physical top quark mass as an eigenvalue that is ∼ m0 %=M& % ≈ 600 GeV. The top quark mass can be adjusted to its experimental value. The diagonalization of the fermionic mass matrix does not aQect the VEV (vweak = 175 GeV) of the composite Higgs doublet. Indeed, the Pagels–Stokar formula is now modi<ed to read: m2t M2 Nc 2 vweak log ≡ f2 = + k ; (4.170) 162 sin2 'L MU 2 where mt is the physical top mass, and 'R the mass matrix diagonalizing mixing angle. The Pagels– Stokar formula diQers from that obtained (in large Nc approximation) for top quark condensation models by the large enhancement factor 1=sin2 'L . This is a direct consequence of the seesaw mechanism. The Top Seesaw employs L = (tL ; bL ) as the source of the weak I = 1=2 quantum number of the composite Higgs boson, and thus the origin of the EWSB vacuum condensate. This neatly separates [258] the problem of EWSB from that of the new weak-isosinglet states in the &L; R and tR sector, a distinct advantage since the electroweak constraints on new physics are not very restrictive of isosinglets. 4.4.2. Dynamical issues How does Topcolor produce the % mass term? We introduce an embedding of QCD into the gauge groups SU (3)1 × SU (3)2 , with coupling constants h1 and h2 , respectively. These symmetry groups are broken down to SU (3)QCD at a high mass scale M . The assignment of the elementary fermions to representations under the full set of gauge groups SU (3)1 × SU (3)2 × SU (2)W × U (1)Y is as follows: L : (3; 1; 2; + 1=3); &R : (3; 1; 1; + 4=3); tR ; &L : (1; 3; 1; + 4=3) : (4.171) This set of fermions is incomplete: the representation speci<ed has [SU (3)1 ]3 , [SU (3)2 ]3 , and U (1)Y [SU (3)1; 2 ]2 gauge anomalies. These anomalies will be canceled by fermions associated with either the dynamical breaking of SU (3)1 × SU (3)2 , or with the b-quark mass generation (a speci SU (3)  2  : t    b   R    (&)L    ::: 354 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 The dynamics of EWSB and top-quark mass generation will not depend on the details of the additional fermions. U 3; 1; 0), which develops a diagonal VEV, We introduce a scalar Higgs <eld, ., transforming as (3; i i .j = V-j . This <eld is presumably yet another dynamical condensate, or it can be interpreted as a relic of compacti (4.172) yielding massless gluons and an octet of degenerate colorons with mass M given by M 2 = (h21 + h22 ) V2 : (4.173) We can also exploit . to provide the requisite M& by introducing a Yukawa coupling of the fermions &L; R of the form: −) &R .&L + h:c: → −M& &&. U We emphasize that this is an electroweak singlet mass term. ) can be a perturbative coupling constant so V)V = M& . Finally, since both tR and &L carry identical Topcolor and U (1)Y quantum numbers we are free to include the explicit mass term, also an electroweak singlet, of the form m0 &L tR + h:c: The Lagrangian of the model at scales below the coloron mass is SU (3)C × SU (2)W × U (1) invariant and becomes L0 = Lkinetic − (M& &L &R + m0 &L tR + h:c:) + Lint : (4.174) Lint contains the residual Topcolor interactions from the exchange of the massive colorons: A h21 MA % M &R + LL + RR + · · · ; (4.175) & R % Lint = − 2 L L M 2 2 where LL (RR) refers to purely left-handed (right-handed) current–current interactions. It suVces to retain in the low energy theory only the eQects of the operators shown in Eq. (4.175) even though higher dimension operators may be present. To leading order in 1=Nc , upon performing the familiar Fierz rearrangement, we have Lint = h21 ( M2 L &R ) (&R L) : (4.176) It is convenient to pass to a mass eigenbasis with the following rede tR → cos 'R tR + sin 'R &R ; (4.177) where tan 'R = m0 : M& (4.178) In this basis, the NJL Lagrangian takes the form: L0 = Lkinetic − MU &R &L + h:c: + h21 [ L (cos 'R &R − sin 'R tR )][(cos 'R &R − sin 'R tR ) L ] ; M2 (4.179) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 µ1 µ1 χ t µ t µ2 t χ + tL Χ t R L + t µ2 (t, χ ) µ1 + µ2 t µ2 χ t ΧR tL (χ ,t) t t µ1 µ1 ΧR tL tL ΧR χ 1 + tL µ1 t t L ΧR t µ2 µ2 + t ΧR t 355 µ2 t µ2 t Χ R Fig. 65. Gap equations for %1 . A similar set of terms is obtained for %2 [468]. where MU = M&2 + m20 : (4.180) At this stage we have the choice of using the renormalization group, or looking at the mass gap equations for %. A rationale for studying the gap equations is that they in principle allow one to explore limits, such as MU ¿ M which are conceptually more diVcult with the renormalization group. 35 For a more complete analysis of the dynamics see [468]. Let us summarize the gap equation analysis. We assume two dynamical mass terms: − %1 tUL &R − %2 tUL tR (4.181) and we compute the gap equations to order O(%3 ). This will produce no IR divergences in terms to order cos2 'R , but there is an IR log-divergence in the last term of Fig. 65 of order sin2 'R , and we take this to be ln(m2t ). Fig. 65 produces the coupled system of gap equations for %1 : 2 h21 N M + MU 2 2 2 2 U %1 = 2 2 %1 cos 'R M − M ln 8 M MU 2 2 2 h21 N %12 M 2 M + MU 2 M + MU 2 2 2 2 + 2 2 %1 cos 'R −%1 ln + 2 − %2 ln 8 M MU 2 MU 2 M + MU 2 2 2 h21 N M + MU 2 M 2 2 2 + 2 2 %2 cos 'R sin 'R M − %1 ln − %2 ln (4.182) 2 U 8 M m2t M 35 For instance, the d = 6 operator makes no sense above the scale M in the renormalization group, but the cut-oQ theory can still be expressed in the gap equation language. 356 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 4 3.5 2.5 eff κc / κc 3 2 1.5 1 0.5 0 0.1 0.2 0.3 0.4 0.5 0.6 0.7 0.8 0.9 1 M/Λ Fig. 66. Gap equation with coupling constant N = h21 =4 (scaled by constant Nc ≡ 2=3) as a function of MU =M , for tan 'R = (0; 0:1; 0:2; 0:3; 0:5; 0:8). From Hill et al. [468]. and a similar set for %2 . With the substitutions %1 = % cos 'R and %2 = % sin 'R the two independent gap equations reduce to a single mass gap equation: 2 h2 N M + MU 2 % = 12 2 % M 2 − MU 2 cos2 'R ln 8 M MU 2 2 2 2 %2 M 2 M + MU 2 M (M + MU 2 ) 2 2 4 −% ln + 2 : (4.183) − % sin 'R ln M + MU 2 m2t MU 2 MU 2 The gap equation (4.183) shows that we require supercritical coupling as the mass MU becomes large. Moreover, for <xed supercritical coupling, h21 =4, as we raise the scale MU the condensate turns oQ like a second order phase transition (Figs. 66 and 67). We can also see that these reproduce normal top condensation in the decoupling limit. For example, choose MU → ∞ for <xed M , and we (4.185) which is the top condensation gap equation. Here we have decoupled &L and &R with MU → ∞. We can also obtain top condensation by setting sin2 'R = 0 and MU → 0, which decouples &L and tR , and causes &R to play the role of tR . A separate question of interest is the structure of the electroweak corrections in the Top Seesaw theory. This is most easily determined from an eQective Lagrangian for the composite Higgs boson. C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 357 1 0.9 0.8 0.7 mtχ /Λ 0.6 0.5 0.4 0.3 0.2 0.1 0 0 0.1 0.2 0.3 0.4 0.5 0.6 0.7 0.8 0.9 1 M/Λ 2 U Fig. 67. Behavior of gap equation for <xed supercritical coupling h1 =4Nc = 2:0 as M =M is increased. In the notation of the The most complete study to date is that of Hill et al. [468]. The composite Higgs boson mass satis<es the approximate NJL result, mHiggs ≈ 2% ∼ 1:2 TeV. Experience with Topcolor suggests that radiative corrections from SU (2)2 will reduce this. The composite Higgs mass is consistent with the unitarity limit of the Standard Model. By itself, this pulls the theory to large negative T . However, there are &–t mixing corrections to the T parameter as well. We obtain for T : 2 4 Nm2t % MU 2%2 T= : (4.186) + 2 ln 2 2 − 1 2 322 vweak Z m20 % m0 MU Putting in typical numbers, such as % = 0:6 TeV, m0 = mt MU =% ≈ 1:2 TeV, Z = 1=128, one observes a large positive correction to T , and one recovers consistency with the error ellipse constraint for MU ≈ 4 TeV. Note that the S-parameter is vanishing, since the additional & quarks are weak isosinglets. We note that, using the freedom to adjust sin 'R , we can in principle dynamically accommodate any fermion mass lighter than ∼ 600 GeV—at the price of some !R : (3; 1; 1; −2=3) : (4.187) 358 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 These fermion gauge assignments cancel the anomalies noted above. We further allow !L !R and !L bR mass terms, in direct analogy to the & and t mass terms: L0 ⊃ −(M! !L !R + m! !L bR + h:c:) : (4.188) With the previous assignments for the & quarks, schematically, this looks like the following: Inclusion of b-quark I SU (3)1   t 1 I=2   b    L    &   I =0 ! R SU (3)  2  : t I =0   b    R     &  I =0 ! L Alternatively, we may de (&; !)R ; : (3; 1; 2; 1=3); (&; !)L : (3; 1; 2; 1=3) : (4.189) Inclusion of b-quark II SU (3)1   t I = 12    b    L    &   1  I=2 ! R SU (3)  2  : t I =0    b    R    &   1  I=2 ! L The schematic model aQords a simple way to suppress the formation of a b-quark mass comparable to the top quark mass. We can suppress the formation of the !L bR condensate altogether by choosing 2 + %2 ∼ M . In this limit we do not produce a b-quark mass. However, by allowing MU ! = %!! !b %!! 6 M and %!b =%!! 1 we can form an acceptable b-quark mass in the presence of a small !L bR condensate. Yet another possibility arises, one which seems to be phenomenologically favored [468], which is to exploit instantons. If we suppress the formation of the !L bR condensate by choosing MU ! ∼ M , there will be a !L bR condensate induced via the ’t Hooft determinant when the t and & are integrated out. We then estimate the scale of the induced !L bR mass term to be about ∼ 20 GeV, and the b-quark mass then emerges as ∼ 20%!b =%!! GeV. We will not further elaborate the b-quark mass in the present discussion, since its precise origin depends critically upon the structure of the complete theory including all light quarks and leptons. Including partners for the b-quark, the T parameter is given by a more general formula [464,468]. The Top Seesaw Model oQers new possibilities for a dynamical scheme explaining both EWSB and the origin of Iavor masses and mixing angles. We focused here in some detail on the third generation and the EWSB dynamics. A fully extended model for light quark and lepton masses has not yet been analyzed, but it would seem that the Top Seesaw aQords interesting new directions C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 359 and possibilities that should be examined. Some varying attempts in this direction can be found in [466,465]. Remarkably, the vectorlike fermions introduced here to provide the seesaw can also help to remedy the discrepancies between lepton and b-quark forward–backward and left–right asymmetries in the LEP data [23]. 4.5. Top Seesaw phenomenology 4.5.1. Seesaw quarks In the Top Seesaw and related models, the third generation quarks [245,246,464], and possibly all Standard Model fermions [466,258], acquire mass through seesaw mixing with exotic, weak-singlet fermions. As a result, there are several new types of states which can aQect the phenomenology: the new fermions themselves, and, composite scalars formed at least in part from the new fermions. We begin by surveying current bounds on the mixing angles 'f between ordinary and weak-singlet fermions. We then look at low-energy limits on the masses of the heavy fermionic seesaw partner, fH , states and the composite scalars (these results apply generally to Kaluza–Klein modes as well). We comment on variant models in which the mixing is with weak-doublet fermions instead of weak-singlets. And to conclude, we discuss Tevatron limits on the masses of the fH states. Suppose the mass matrix of a Standard Model top quark mixing with a vectorial isosinglet quark & has the seesaw form of Eq. (4.165). The matrix is diagonalized by performing separate rotations on the left-handed and right-handed fermion <elds (Eq. (4.166)). Most of the phenomenology is sensitive to the mixing among the left-handed fermions. Of the two mass eigenstates, the physical top quark is the lighter one t˜L and is mostly weak-isodoublet: t˜L = cos 't tL − sin 't &L ; (4.190) ˜ is mostly weak-isosinglet: and has a mass of order m0 %=M& ; the heavier (&) TLH ≡ &˜L = sin 't tL + cos 't &L (4.191) with a mass of order ∼ M& . This structure is readily generalized to models in which more than one ordinary fermion mixes with weak singlets [469]. Recent limits on the mixing angles between ordinary and weak-isosinglet fermions from precision electroweak data were obtained in [470]. Separate limits at 95% c.l. are given for the case in which each fermion Iavor has its own weak partner [466,258] sin2 'e 6 0:0024; sin2 '% 6 0:0030; sin2 'd 6 0:015; sin2 's 6 0:015; sin2 'u 6 0:013; sin2 'c 6 0:020 sin2 '= 6 0:0030 ; sin2 'b 6 0:0025 ; (4.192) (4.193) (4.194) and the case in which only the third-generation quarks have weak partners [245,246,464] sin2 'b 6 0:0013 : (4.195) The most important precision electroweak constraints on the Top Seesaw come from the S and T parameters [464,470,468]. It is remarkable that the minimal Top Seesaw model, which typically includes a heavy composite Higgs boson around 1 TeV (see Fig. 68), is non-trivially compatible with 360 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 1 6 5 Mh (TeV) 10 2 Topseesaw Higgs Mass at Large-Nc: exact Mh (solid) & Mh≈2mtχ (dotted) 4 (a) 10 3 2 1 0.9 0.8 0.7 Bottom to top: κ/κc = 1.05, 1.2, 1.5, 2, 4 1.2 Higgs Mass by Improved RG Mh (TeV) 1 Curves from left to right: κ/κc = 4, 2, 1.5, 1.2, 1.05 0.8 0.6 0.4 0.2 1 (b) 10 10 2 Λ (TeV) Fig. 68. The predicted mass spectrum of the Top Seesaw Higgs boson: (a) by the large-Nc fermion-bubble calculation; and (b) by an improved RG analysis including the Higgs self-coupling evolution. the S–T bounds. The composite Higgs boson’s contributions will drive T in the negative direction relative to a light SM Higgs. However, the Top Seesaw sector has generic weak-isospin violation from the t–& mixing which will signi C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 -0.4 -0.2 0 0.2 -0.4 -0.2 0 0.2 5.0 (b) 0.8 0.8 κ/κc = 1.05 0.6 5.0 TeV 3.3 TeV 95%C.L. 3.8 4.3 0.4 κ/κc = 1.2 (a) 3.8 95%C.L. 0.1 0.2 6.3 ✣ 4.9 8.7 7.5 13 0.2 13 = Mχ 0.5 0 0.4 6.5 = Mχ 0.3 0.6 6.3 7.5 5.4 ✣ ∆T 361 0 mSM h =1.0 TeV Right curve: Fermion-bubble Left curve: Improved RG 0.8 0.8 10 9.8 (c) 24 0.6 10 95%C.L. 13 0.4 95%C.L. 13 16 ✣ ✣ 26 0.2 (d) 22 TeV 24 0.4 36 30 48 36 = Mχ 60 16 TeV = Mχ 0.6 0 0.2 0 κ/κc = 2 κ/κc = 4 -0.2 -0.2 -0.4 -0.2 0 0.2 -0.4 ∆S -0.2 0 0.2 Fig. 69. Top Seesaw contributions to S and T are compared with the 95% c.l. error ellipse (with mref h = 1 TeV) for N=Nc = 1:05; 1:2; 2; 4, shown as a function of M& . In each plot, the curve on the right is derived from the large-Nc fermion bubble calculation, the curve on the left is deduced by an improved RG approach. For reference, the SM Higgs SM corrections to (S; T ), relative to mref h =1 TeV, are given for mh varying from 100 GeV up to 1:0 TeV in plot (a) (from Hill, He and Tait [468]). precision data are eQectively probing M& ∼ 4 TeV. In Fig. 70, we display the same S–T trajectories as in Fig. 69, but with the corresponding Higgs mass (Mh ) values marked [468]. The inclusion of a bottom seesaw generates additional b–! mixing (! ≡ bH is the seesaw partner U vertex of b) which makes nontrivial contributions to the S and T parameters and also to the Z → bb (Fig. 71). Furthermore, the composite Higgs sector now contains two doublets and thus provides additional corrections to the precision observables. The b–! mixing induces a positive shift in the left-handed Z–b–bU coupling e -gLb = + (sin 'b )2 ; (4.196) 2 sin ,W cos ,W b 2 U which results in a decrease of Rb = E[Z → bb]=E[Z → hadrons], i.e., Rb RSM b − 0:39(sL ) , as obtained in Refs. [464,468]. This puts an upper bound on the b-seesaw angle, mb =%! mb sin 'b √ √ ; M ! rb 1 + rb (4.197) 362 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 -0.4 -0.2 0 0.2 -0.4 -0.2 0 0.2 0.8 0.8 κ/κc = 1.05 0.6 482 95%C.L. 0.4 κ/κc = 1.2 (a) 972 GeV 95%C.L. 965 471 0.1 0.2 (b) 1.11 TeV 1.1 ✣ 955 427 1.09 .434 0.2 1.05 = Mh 0.5 ∆T 0 0.4 943 = Mh 0.3 0.6 .482 .468 455 ✣ .501 0 mSM h =1.0 TeV Right curve: Fermion-bubble Left curve: Improved RG 0.8 0.8 .489 1.324 (c) .459 0.6 1.320 95%C.L. .471 0.4 95%C.L. .455 0.4 ✣ 1.307 0.6 1.42 .434 ✣ 1.41 .418 .426 0.2 (d) 1.43 TeV .407 1.294 TeV = Mh 1.40 = Mh 0 0.2 0 κ/κc = 2 κ/κc = 4 -0.2 -0.2 -0.4 -0.2 0 0.2 -0.4 ∆S -0.2 0 0.2 Fig. 70. Same as Fig. 69, but with the corresponding Mh values marked on the S–T trajectories instead (from Hill et al. [468]). and correspondingly a lower bound on the mass M! ( M& ), as summarized in Fig. 72. The Rb bound will mainly constrain the low tan region of the eQective composite two doublet model. Variant models [417,467] in which vectorial weak-doublet partners exist for both top and bottom give similar contributions to T as the models with a weak-singlet partner for top, while the contributions to Rb are suppressed by a small mixing angle. The lower bound on these exotic quarks is, then, of order a few TeV, assuming they are degenerate. In principle, one could experimentally distinguish between the models with weak-singlet and weak-doublet mixings by measuring AtLR at an NLC [467]. The predicted shifts relative to the SM value would be of similar size but opposite sign, as the ZtL tL coupling is altered in the weak-singlet models while the ZtR tR coupling is altered in the weak-doublet models. As discussed in [470], it is possible to use existing Tevatron data to set limits on direct production of the &-like states. New, mostly-singlet, quarks decaying via mixing to an ordinary quark plus a W boson would contribute to the dilepton events used by the CDF [471] and D0 [472] experiments to measure the top quark production cross-section. Since the weak-singlet quarks are color triplets, they would be produced with the same cross-section as sequential quarks of identical mass. However the weak-singlet quarks can decay via neutral-currents (e.g. dH → ZdL ) as well as charged-currents (e.g. dH → WuL ), and this lowers the branching fraction of the produced quarks to the C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 -0.4 -0.2 0.2 0 -0.4 -0.2 0.2 0 0.8 0.8 (a) κ/κc = 2, tanβ = 2 0.6 (b) 17 6.3 95%C.L. 0.6 95%C.L. 20 0.4 ✣ 0.4 23 26 TeV = Mχ 25 TeV = Mχ 32 0.3 0.5 0 ✣ 19 0.1 0.2 ∆T 363 0.2 0 mSM h = 1.0 TeV tanβ = 5 0.8 0.8 13 (c) (d) 12 0.6 95%C.L. 95%C.L. 16 0.4 ✣ 13 0.4 ✣ 20 23 TeV = Mχ 0.2 0.6 0.2 20 26TeV = Mχ 0 tanβ = 12 0 tanβ = 40 -0.2 -0.2 -0.4 -0.2 0 0.2 -0.4 ∆S -0.2 0 0.2 Fig. 71. Top and bottom seesaw contributions to S and T are compared with the 95% c.l. error ellipse (with mref h = 1 TeV) for N=Nc = 2 and & = 3 × 10−3 with a variety of values of tan . The S–T trajectories (including both Higgs and quark contributions) are shown as a function of M& . For reference, the SM Higgs corrections to (S; T ), relative to mref h = 1 TeV, are depicted for mSM varying from 100 GeV up to 1:0 TeV in plot (a) (Ref. [468]). h to which the search is sensitive. In fact, the decay width bH → cL W is so strongly suppressed both by Cabbibo factors and the large rate of bH → bL Z, that the Tevatron data do not provide a lower bound on MbH . For models in which all quarks have weak-singlet partners, the limits 153 GeV CDF Md;Hs ¿ : (4.198) 143 GeV D0 Similarly, mass limits on new mostly-singlet leptons (for Iavor-universal mixing models) can be extracted from the results of LEP II searches for new sequential lepton doublets. In the relevant searches, the new neutral lepton N is assumed to be heavier than its charged partner L and L is assumed to decay only via charged-current mixing with a Standard Model lepton (i.e. BR(L → (‘ W ∗ ) = 1:0). The OPAL [473] and DELPHI [474] experiments have each set a 95% c.l. lower bound of order 80 GeV on the mass of a sequential charged lepton. As detailed in [470], when one adjusts for the increased production rate and decreased charged-current branching fraction of the 364 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 1 10 10 2 1 κ/κc = 2 tanβ = 1 -1 3 sinθLb 10 ⇓ 10 5 Allowed -2 2σ Rb bound 12 40 (a) 10 -3 10 2 min Mχ,ω (TeV) (tanβ, Mχ,ω ) = (1, 20) (3, 9) 10 (5, 5.5) (12, 2.3) (Different tanβ curves largely overlap) min (Mχ,ω in TeV) 1 1 (b) 10 10 2 Λ (TeV) Fig. 72. The Rb limits are shown for the b seesaw angle sLb = sin ,Lb in plot (a) and for the mass M! ( M& ) in plot (b). Here, we choose N=Nc = 2 and a wide range of tan values. mostly-singlet ‘H states, the resulting limit is M‘H ¿ 84:9 GeV (4.199) is slightly stronger. Finally, CDF limits [475] on heavy bH quarks pair-produced through QCD processes and decaying via neutral currents can be applied to the weak-singlet fermions so long as the value of B(bH → bL Z 0 ) is included. CDF C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 σ ⋅BR(b′ → bZ)2 (pb) 10 365 2 – – Theory σ(pp→b′b′X) 95% CL limit 2 199 GeV/c 10 1 100 120 140 160 180 b′ mass (GeV/c 2) 200 220 Fig. 73. The 95% c.l. upper limit on ppU → b bU X production cross section times the b → bZ branching ratio squared (solid). The dashed curve shows the value predicted in a theory where the branching ratio is 100% [475]. 10 4 10 mχ = 1 TeV 3 mχ = 3 TeV σχχ (fb) 10 2 10 mχ = 5 TeV 1 10 10 10 10 -1 MG = 10 TeV -2 p pbar -3 pp -4 20 40 60 80 100 120 140 E (TeV) Fig. 74. Cross-section for pair production of & states at high energy hadron colliders, leading to 6-top events [194]. 4.5.2. Flavorons As an extension of the Top Seesaw mechanism, models with Iavor universal dynamical symmetry breaking have been proposed [466,476]. In these models, the dynamics are driven by family or large Iavor gauge symmetries; when these symmetries break, multiplets of heavy “Iavoron” bosons remain in the spectrum. Their masses, like the symmetry breaking scale, are expected to be of order a few TeV. The Iavorons’ couplings to fermions must be strong enough to generate condensates that break the electroweak symmetry and provide the ordinary fermions with masses. As a result, the Iavorons tend to be readily produced at hadron colliders like the Tevatron—but also to have large decay widths that make their detection in dijet invariant mass spectra a challenge [466]. 366 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 For the moment, some of the best limits on Iavor gauge bosons come from precision electroweak data [454]. Flavorons coupling to ordinary fermions can give rise to direct corrections to Zff vertices; generally speaking, if the Zff coupling is gf , the correction is [429] 2 GNF MZ2 MF ; (4.200) ln Zgf = gf 2 6 MF MZ2 where G, N ≡ gF2 =4, and MF are the group theory factor, coupling, and mass appropriate to the particular Iavoron. Flavoron exchange across t and b quark loops will generally contribute to the T parameter an amount (in leading log) T = Nc (GLL + 2GRR ) Nm4t [ln(MF2 =m2t )]2 ; 322 sin2 ,W cos2 ,W MZ2 MF2 (4.201) where GLL; RR are group theory factors for Iavoron couplings to left-handed and right-handed quarks. In addition, in models in which some generator of the Iavor group mixes with hypercharge or the diagonal generator of SU (2)W , the associated Z − Z mixing also contributes to T . For each of the three representative Iavoron models mentioned here, we assume that Iavoron exchange at low energies may be treated in the NJL approximation (with four-fermion coupling 4N=2MF2 ). The critical coupling required for chiral symmetry breaking is 2N : (4.202) Ncrit = 2 N −1 The minimal model [466,476,477,83] has a gauged SU (3) family symmetry acting on the left-handed quark doublets. This family symmetry group does not mix with the Standard Model gauge groups and the model leaves the GIM mechanism intact. The critical coupling is Ncrit = 2:36. Models in which the SU (9) symmetry of color and family multiplicity of the left-handed quarks is gauged [466,265] also preserve GIM. A proto-color group acting on the right-handed quarks is also present; obtaining the correct value for s in its presence requires N ¿ 3 s (2 TeV) ≈ 0:3 : (4.203) Ignoring mixing between the Iavor and proto-color group generators, the critical coupling for chiral symmetry breaking is Ncrit = 0:71. If the full SU (12) Iavor symmetry of all the left-handed quark and lepton doublets is gauged [466,265], both a proto-color group and a proto-hypercharge group must be included for the right-handed fermions, and the constraint (4.203) still applies. In this case, Ncrit = 0:53—not far above the lower limit. The 95% c.l. lower bound on MF from precision electroweak data for N at its critical value is approximately 2 TeV for all three models [454]. Comparable limits on MF (for critical coupling N) in the SU (3) and SU (9) models have been obtained from Tevatron Run I data [478]. Run II should be sensitive to MF up to 2.5 –3 TeV with 2 fb−1 of data. For the SU (3) model, both dijet and anomalous single top production are likely to yield large signals; in the SU (9) model, the dijet signal is relatively enhanced and the single top signal, relatively reduced in size, enabling the two models to be distinguished if both channels are studied. A signi C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 367 4.6. Extra dimensions at the TeV scale Recently there has been considerable interest in theories of extra dimensions of space which emerge not far from the weak scale. We cannot give a comprehensive review of this burgeoning <eld, but we will outline some key ideas relating to electroweak symmetry breaking and the physics of new strong dynamics. The ideas largely stem from the view that string theories admit an arbitrary hierarchy between the compacti Possibly the earliest proposal for “weak scale extra dimensions” was an heroic eQort by Darrell R. Jackson to identify the W and Z as Kaluza–Klein modes (private communication to CTH, Caltech, ca. 1975). 37 For a discussion of one possibility for extra dimensions of time, see [495]. 368 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 The idea of extra c-number dimensions near the weak scale has led to the proposal of dynamical schemes that can imitate the strong dynamics we have discussed in this review. Dobrescu (4.204) we can arrange for . to develop a vacuum expectation value V , by which all unwanted Higgs degrees of freedom are elevated to the large mass scale ˙ M . Note that in trying to discard the extra 10 degrees of freedom in . as a Higgs <eld, we encounter a unitarity bound [24,506,507] which predicts the string scale from M . The Lagrangian containing the gluon and the (4.205) The <eld ., after developing a VEV, and decoupling the Higgs degrees of freedom, can be parameterized by a chiral <eld: √ . → V exp(i'A T A = 2V ) : (4.206) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 369 This is exactly the structure that would emerge from a Wilson lattice with two transverse lattice points (which are now “branes, each carrying a 3+1 theory). We also see that this is identical to the gauge sector of Topcolor models. With the breaking in place, the theory has two mass eigenstates, the gluon and coloron: G%A = cos ,AA1% + sin ,AA2% ; B%A = −sin ,AA1% + cos ,AA2% ; (4.207) where Ai% belongs to SU (3)i . If we allow SU (3)i to have a coupling constant gi then we tan , = (4.208) The mass of the coloron is given by 1 MK = V g12 + g22 = ; R (4.209) where R is identi<ed with the radius of the compacti<ed extra dimension. We also see that the low energy QCD-coupling is related to gi as 1 2 gQCD = 1 1 + 2 : 2 g1 g2 (4.210) 2 With g1 =g2 we have g12 =2gQCD and this is the onset of “power-law running” of the coupling constant. The model indeed represents a single KK mode, descending from higher dimensions. Conversely, observation of a coloron in the particle spectrum could be taken to suggest that a new extra dimension is opening up. The full lattice description of QCD is a straightforward generalization of this model [41]. If we consider the full lattice gauge theory in 4 + 1 dimensions with a Wilson action for the extra dimension, we obtain the Lagrangian: N −1 N L=− 1 A A%( F F + Tr(D% .k )2 : 4 j=1 j%( j (4.211) k=1 Here there are N − 1 . <elds, and the kth <eld transforms as an (3k ; 3Uk+1 ) representation, straddling the nearest neighbor SU (3)k and SU (3)k+1 gauge groups. .k is a unitary matrix, and represents the Wilson line: xk+1 (4.212) .k = V exp ig d x5 Aa5 T a : xk .k may be parameterized by √ .k = V exp(i'ak T a = 2V ) ; where T a = 1a =2. This clearly matches our previous discussion of a single KK mode. (4.213) 370 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 The .k kinetic terms imply a gauge <eld mass term of the form: N −1 1 2 2 a gV (Ak% − Aa(k+1)% )2 : 2 (4.214) k=1 Performing the mass spectrum analysis of the model is equivalent to analyzing the phonon spectrum of a one-dimensional crystal lattice. The mass term is readily diagonalized and we obtain eigenvalues reIecting a ladder of states with √ Mn = 2gV sin(n=2N ) (4.215) the lowest energy masses given by √ 1 : (4.216) Mn ≈ gVn=N 2 ≡ R This is the spectrum assuming free boundary conditions on the lattice and corresponds to orbifolding. With periodic boundary conditions instead, one gets the Aa5 as low energy modes as well, which act like pseudo-scalars, and the KK mode levels are doubled. The net result is that the eQective Lagrangian for, e.g., QCD propagating in the extra dimensional bulk, is a chain of gauge groups of the form: SU (3) × SU (3) × SU (3) × · · ·, one √ gauge group per lattice brane. The high energy coupling constant common to each group is g ∼ N gQCD , consistent with the power-law running. From a “bottom-up” perspective, one writes down this gauge structure with arbitrary relevant operators with arbitrary coeVcients to describe the most general low energy eQective theory for extra-dimensional physics. Various additional dynamical issues in deconstructed theories are currently under investigation, such as chiral dynamics [509,510], electroweak observables [511,512], uni If the background <eld is approximately constant both chiral components of the fermion appear on each lattice brane, as depicted in Fig. 75. The cross bars are the latticized links allowing hopping of the fermion from n to n + 1 in Eq. (4.217). If '(x5 ) swings through zero rapidly in the vicinity of brane n, i.e., 'n = 0 then only a single chiral component is normalizable in the vicinity of n and one gets a dislocation in the lattice as shown in Fig. 76. The coupling strength of SU (3)n on the nth brane will generally be renormalized by the dislocation and can become supercritical, Fig. 77. It would, therefore, not be coincidental to expect this to happen; indeed a variety of eQects are expected to occur near the dislocation, e.g., the chiral fermions themselves can feed-back onto the gauge <elds to produce such renormalization eQects. The result C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 1 2 ... n n+1 ... 371 N ΨR X X X X X X X ΨL Fig. 75. Dirac fermion corresponding to constant ' has both chiral modes on all branes. The ×’s denote the ' couplings on each brane, and the links are the latticized fermion kinetic terms which become Wilson links when gauge <elds are present (from [508]). The spectrum here has a singlet lowest massive mode, and doubled KK modes; by adding a Wilson term one can remove one of the two cross-bars between adjacent branes, and eliminate second Brillouin zone doubling in the spectrum. 1 2 ... n n+1 ... N ΨR X X X X X X ΨL Fig. 76. A chiral fermion occurs on brane n where '(x5 ) swings rapidly through zero. The chiral fermion has kinetic term (Wilson links) connecting to adjoining branes (from [508]). 372 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 1 2 ... n n+1 ... N ΨR X X X X X X X X ΨL tR X X X X tL Fig. 77. Pure top quark condensation by Topcolor is obtained in the limit of critical coupling on brane n and decoupling to the nearest neighbors. Decoupling corresponds to taking the compacti is a chiral condensate on brane n forming between chiral fermions. Identify 3 = (t; b)L and tR as the chiral zero-modes on brane n of two independent Dirac <elds in the bulk. In the limit that we take the compact extra dimension very small, the nearest neighbor links decouple at low energies. In this limit we recover a Topcolor model with pure top quark condensation. In Fig. 78 we illustrate the case that some of the links to nearest neighbors are not completely decoupled. Again, this can arise from renormalizations due to background <elds, or to warping [503,508]. Thus the mixing with heavy vectorlike fermions occurs in addition to the chiral dynamics on brane n. In this limit we obtain the Top Quark Seesaw Model. 4.6.2. Little Higgs theories Recently another approach to electroweak symmetry breaking has been revived, inspired by deconstruction, in which the Higgs itself is a PNGB [501]. This is actually an old idea, due to Georgi, Kaplan, and others (see, e.g. [525–527]). The novelty presently is that delocalization in an extra dimension (or the minimal deconstructed description of such) leads to a potentially softer scheme in which dangerous gauge loop quadratic divergences contributing to the Higgs boson (mass)2 can be cancelled in one loop order. The resulting models bear some resemblance to the MSSM Higgs structure. There are now a number of schemes, most notably the “minimal moose model”, [528] and SU (5)=SO(5) model [529,530], and the SU (6)=SO(6) scheme [531]. We brieIy describe the “minimal moose” scheme. When we periodically compactify a higher dimension we C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 1 2 ... n n+1 ... 373 N ΨR X X X X X X X X ΨL tR X X X X tL Fig. 78. Top Seesaw Model arises when the eQects of nearest neighbor vectorlike fermions are retained, i.e., when these heavier states are only partially decoupled. Keeping more links maintains the seesaw. Usually we denote tRn ∼ &R , tLn+1 ∼ &L , tRn+1 ∼ tR (from [508]). a Coleman–Weinberg potential for the modes, and is generally 374 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 5. Outlook and conclusions The next decade will bring great discoveries in the <eld of elementary particle physics. The Tevatron Run-II program and the LHC will begin to reveal the mechanism(s) responsible for the origin of mass. The Large Hadron Collider in concert with a Linear Collider can begin detailed studies of the <elds involved in electroweak symmetry breaking. In the longer term, a Very Large Hadron Collider will have the power to probe the corollary mechanisms that underlie the physics of Iavor. This review has focused on the possibility that the origin of mass involves a new strong dynamics near the TeV scale. As we have argued here, dynamical electroweak symmetry breaking, whether it comes in the guise of Technicolor, Topcolor, or some variant scheme, can provide a natural explanation for the weak scale. It can also provide tantalizing clues about the magnitudes and origins of the fermion masses and mixings, and may be able to explain CP violation. Indeed, many issues of Iavor physics that are relegated to the Planck or Gut scale in perturbative theories, such as the MSSM, become accessible to suVciently energetic accelerators in the dynamical framework. Technicolor was created in the era in which the W and Z bosons are heavy and all known fermions are comparatively light. In the C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 375 generate all known fermion masses and mixing angles, though no complete theory has yet been written down. Thus, strong dynamical models of EWSB have evolved and remain viable and consistent with all known experimental limits. Of course, regardless of the theoretical beauty of any given scenario, its validity will ultimately be ascertained through experiment. Should the slight discrepancies in the electroweak precision tests, e.g., the forward–backward asymmetries of leptons and hadrons, open into a new phenomenology, then novel TeV-scale dynamics may be in the oVng. Should the hadron colliders 376 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 fermion mass terms are not singlets under the gauge group, owing to the fermionic chiral charge assignments. The gauge <elds also acquire their masses through the condensing Higgs boson, as per the toy example of Section 1.3(iv). We begin by considering an SU (2) × U (1) gauge theory with one scalar <eld that transforms as an isodoublet under the SU (2) group. The covariant derivative for the theory is given by Y iD% = i9% − g2 W%a Qa − g1 B% 2 Y ; (A.1) = i9% − g2 W%+ Q− − g2 W%− Q+ − g2 W%3 Q3 − g1 B% 2 √ √ where the “charge basis” expressions are Q± = (Q1 ± iQ2 )= 2, and W%± = (W%1 ± iW%2 )= 2. Note there are two gauge coupling constants, g2 and g1 , one for each simple subgroup of the theory. The Qa are the SU (2) weak charges and Y is the U (1) hypercharge. The Qa satisfy the SU (2) Lie algebra: [Qa ; Qb ] = iHabc Qc : (A.2) These charges are at this stage abstract operators, the generators of the associated groups. We specialize to particular representations for the charges, choosing an I = 12 weak isospin representation for the left-handed fermions, and Higgs boson, in which Qa = = a =2 where the = a are Pauli matrices. The right-handed fermions are singlets, annihilated by the Qa . We further de (A.6) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 377 where Z%0 (A% ) is the physical Z-boson (photon), then we see, g2 sin , = e; g1 cos , = e : (A.7) The photon thus couples to eQEM with strength e where: 1 1 1 = 2+ 2 : 2 e g 2 g1 The weak mixing angle is de (A.8) (A.9) The gauge covariant <eld strengths are de i Tr(= a [D% ; D( ]) = 9% Aa( − 9( Aa% + Habc Ab% Ac( ; g22 F%( = − i −1 Y Tr([D% ; D( ]) = 9% B( − 9( B% : g12 r (A.10) Then the gauge <eld kinetic terms are LG:B: kinetic 1 a a 1 = − F%( F%( − F%( F%( : 4 4 Consider the complex doublet scalar Higgs-boson with Yr = −1: 0 ' : H= '− (A.11) (A.12) The Lagrangian for H takes the form: L = (D% H )† (D% H ) − V (H ) : (A.13) The masses of the gauge bosons are generated by spontaneous symmetry breaking, in strict analogy to the generation of mass in a superconductor as discussed in Section 1.3(iv). We assume that V (H ) has an unstable extremum for H = 0 and a nontrivial minimum, e.g., we can write: V (H ) = 1 2 (H † H − vweak )2 : 2 (A.14) The Higgs boson then develops a vacuum expectation value. This may, without loss of generality be taken in the upper component (which is, again, the choice of orientation within the internal SU (2) space that de 378 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 When we substitute the ansatz (A.15) into the Higgs boson kinetic term of Eq. (A.13), the masses of the gauge bosons are generated (as in the Abelian Higgs model of Section 1.3(iv)): √ 1 1 1 2 + %− 1 2 1 2 2 1 2 % L = (9h) + MW W% W + MZ Z% Z − MH h − MH h3 − 1h4 2 2 2 2 2 8 1 MH h2 + h (g22 W%+ W %− + (g12 + g22 )Z% Z % ) ; + (A.16) 2 1 √ where MH = vweak 21. This Lagrangian also exhibits the coupling of the Higgs h <eld to itself and to the physical W and Z <elds. Inverting Eq. (A.6): A% = sin ,W%3 + cos ,B% ; Z% = cos ,W%3 − sin ,B% = (g2 W%3 − g1 B% ) g12 + g22 allows us to rewrite the covariant derivative as Y + − − + 3 − Z% Q˜ ; iD% = i9% − g2 W% Q − g2 W% Q − eA% Q + 2 where the Z-boson couples to the neutral current charge: =3 Y Q˜ = g12 + g22 cos2 , − sin2 , 2 2 =3 Y = e cot , − tan , : 2 2 (A.17) (A.18) (A.19) We can also extract the masses, upon comparing Eq. (A.16) with Eq. (A.13) using the full covariant derivative (A.18): MW2 = 1 2 2 ; gv 2 2 weak MZ2 = 1 2 (g2 + g22 ) : v 2 weak 1 (A.20) From this, it follows: MW2 = cos2 ,W ; MZ2 1 2 vweak = √ : 2 2GF (A.21) The coupling of the matter <elds to gauge <elds is then determined through the gauge-invariant kinetic terms. For example, let us consider the left-handed top and bottom quark doublet 3L =(t; b)L . The kinetic term is: 1 1 U %− 3U L iD3L = 3U L i3L − √ tU % LbW %+ − √ b % LtW 2 2 − 2e e U U ˜ tU % LtA% + bLbA % − 3L Q % 3L Z% 3 3 (A.22) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 379 where L = 12 (1 − 5 ). Dirac mass terms would be of the form 3U L R and are therefore isospin − 12 ; we cannot directly add such nonsinglet terms into the Lagrangian. We can, however, assume there are terms of the form: gt 3U L · HtR + gb 3U L · H c bR (A.23) which couple the left- and right-handed fermions to the Higgs <eld, and which are invariant under SU (2)×U (1). When H develops its VEV, we see that we obtain masses mt =gt vweak and mb =gb vweak for the top and bottom quarks respectively. In general, of course, there will occur mixing eQects when we include the light generations; these lead to the Cabibbo–Kobayashi–Maskawa (CKM) matrix relating the gauge and mass eigenbases for left-handed quarks. For completeness, we mention that to minimally generate a neutrino Majorana mass, we would write for a leptonic doublet 3L = ((; ‘)L : g( U (3L H )(H C 3LC ) ; M (A.24) 2 where 3LC is the charge-conjugated spinor. This term produces a Majorana mass for (L of g( vweak =M . The interplay between gauge symmetries, and chiral symmetries, both of which are broken spontaneously, is fundamental to the Standard Model. The left-handed fermions carry the electroweak SU (2) quantum numbers, while the right-handed do not. All of the mathematical features of the symmetric Lagrangian remain intact, but the spectrum of the theory does not retain the original obvious symmetry properties. When a massive gauge boson was discovered, such as the W ± or Z of the Standard Model, we also discovered an extra piece of physics: the longitudinal component, i.e., the NGB which comes from the symmetry breaking sector. (b) Summary of one-loop oblique radiative corrections We will now summarize a particular subclass of electroweak radiative corrections, the so-called “oblique” corrections. To begin, we generalize slightly our de MZ2 = 12 vZ2 (g12 + g22 ) ; (A.25) 2 2 =vZ2 =vweak , where vW (vZ ) is the Higgs VEV “as seen by” the W -boson (Z =boson). In tree level vW 2 and we want to include radiative eQects that split these quantities and lead to their q evolution. The radiative corrections to the spontaneously broken Standard Model thus involve keeping track of four 2 underlying functions of the momentum scale, q2 which are: g12 (q2 ), g22 (q2 ) and vW (q2 ) and vZ2 (q2 ). For the running of the coupling constants, to a good approximation, we really only need to know (MZ ) ≈ (MW ) in most applications. Let us be more precise about vW and vZ . First, we scale away the coupling constants by de (A.26) 380 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 (superscript T stands for “transverse”). We then make the speci (A.27) and of the couplings 1 1 1 T T T = 2 − Q33 − Q3B = 2 − Q3Q ; g22 g2un g2un 1 1 1 T T T T = 2 − QBB − Q3B = 2 + Q3Q − QQQ ; 2 g1 g1un g1un (A.28) where gi un is an unrenormalized coupling constant. The dominant Standard Model contributions to 2 vW and vZ2 at q2 = 0 come from the third generation and a putative Higgs boson. The diQerence 2 vW − vZ2 is, upon computing the loops, MZ2 m2H MW2 m2H 2 2 2 2 ln(mH =MW ) − 2 ln(mH =MZ ) : (A.29) + 2 mH − MW2 mH − MZ2 The parameter of Veltman is 2 Nc 2m2t m2b vW 2 2 (m + m ) − = 2 =1+ log(m2t =m2b ) t b 2 2 2 2 vZ 32v0 (mt − mb ) M 2 m2 M 2 m2 + 2 W H 2 ln(m2H =MW2 ) − 2 Z H 2 ln(m2H =MZ2 ) mH − MW mH − MZ : (A.30) 2 In general, vW and vZ2 have small, nonzero, q2 dependence. Moreover, the eQect of new physics 2 at the weak scale, |q| ∼ 1 TeV will generally be to induce additional q2 dependence into vW and 2 2 2 vZ and additional splitting into vW − vZ beyond the Standard Model contributions. For example, new sequentially heavier chiral fermions will contribute through loops, and their contributions do not decouple as their masses are taken large. The squared couplings g12 and g22 are less susceptible to eQects from new physics. We seek, therefore, a convenient parameterization of the eQects of new physics that can be used for comparison with a variety of experiments. The resulting interpolation between diQerent experiments may involve nonstandard values of these new parameters and yield evidence for new physics. 2 A simple set of parameters can be generated by expanding vW and vZ2 in a Taylor series in q2 as follows: 2 2 vW (q2 ) = vweak + q2 + =vr2 + !q2 ; (A.31) 2 + q2 − =vr2 − !q2 : vZ2 (q2 ) = vweak (A.32) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 381 2 Thus, vweak is the average (I = 0) of the two decay constants and contains most of the physics of symmetry breaking, while = (I = 1) is just a rewriting of the splitting between these at zero momentum (i.e., the parameter). The parameters and ! are respectively (I = 0) and (I = 1) measures of physics contributing to the q2 evolution in the eQective low energy theory. A similar parameterization exists in the literature and is due to Kennedy and Lynn, [538], Altarelli and Barbieri [539] and Peskin and Takeuchi [235]. These authors consider the coeVcients of the full vacuum polarization tensors QXY . Peskin and Takeuchi de 9 9 2 2 ; (A.33) Q | − Q | S = 16 33 q =0 3Q q =0 9q2 9q2 4 [QWW |q2 =0 − Q33 |q2 =0 ] ; sin , cos2 ,MZ2 9 9 U = 16 QWW |q2 =0 − 2 Q33 |q2 =0 : 9q2 9q T= 2 (A.34) (A.35) 2 and vZ2 : Recalling our de 1 2 1 2 T v = v + Q33 − Q3Q 2 Z 2 weak (A.36) 1 2 1 2 T vW = vweak + QWW − Q3Q 2 2 (A.37) and we see that = 2S + U ; 16 != U ; 16 == sin2 , cos2 ,MZ2 T = T : 2 4vr 2 (A.38) It is straightforward to compute S, T and U . One Nc [1 − Y log{m2b =m2t }] : 6 (A.39) T is given by the usual -parameter expression, or the diQerence of the zero-momentum expressions 2 for vW and vZ2 : 2m2t m2b Nc 2 2 (m + m ) − T= log(m2t =m2b ) t b (m2t − m2b ) 4 sin2 , cos2 ,MZ2 MZ2 m2H MW2 m2H 2 2 2 2 ln(mH =MW ) − 2 ln(mH =MZ ) : (A.40) + 2 mH − MZ2 mH − MW2 Similarly, for U we can obtain 5m4t − 22m2t m2b + 5m4b m6t − 3m4t m2b − 3m2t m4b + m6b Nc 2 2 − U= + log{mt =mb } : 6 3(m2t − m2b )2 (m2t − m2b )3 (A.41) 382 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 √ Combining the inputs from diQerent measurements, such as GF =1=2 2vweak (0)2 , MZ2 = 12 (g12 (MZ2 )+ 2 2 (MZ2 ), MW2 = 12 g22 (MW2 )vweak (MW2 ), sin2 ,Z -pole = g1 (MZ2 )=(g1 (MZ2 ) + g2 (MZ2 )), and mtop , g22 (MZ2 ))vweak one identi<es (at some chosen con<dence level) an ellipse-shaped region of the S–T plane with which the data is most consistent. The S–T coordinate system is chosen with arbitrary oQsets so that some particular Standard Model value of MH de a L + U aR i9= a R + G( U aL Ub Ra )( R Lb ) ; (B.1) where (a; b) are global SU (N ) “color” indices. Note that the four-fermion form of the interaction is contained in a Fierz rearrangement of a single-coloron (massive gluon) exchange potential: 1A 1 2 U 1A g%( U % ( g (B.2) 2 2 q 2 − 2 2 thus G = g2 =2 at q2 = 0. Thus, in some sense we may view this an approximation to what we believe is happening in QCD on scales M 2 ∼ 2QCD , ignoring the eQects of con RH + h:c:) − 2 H † H : (B.3) If we integrate out the <eld H we reproduce the four-fermion vertex as an induced interaction with G ≡ g2 =2 . Note that G ¿ 0 implies an attractive interaction and permits the factorization in this form. More speci C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 383 Lagrangian on a scale % ¡ in the fermion bubble approximation we integrate out the fermion <eld components on scales % ↔ . The full induced eQective Lagrangian at the new scale % will then take the form: 10 (H † H )2 − )0 RH † H : (B.4) 2 Here R is the geometric scalar curvature, and we see there is an induced “nonminimal” coupling of the Higgs <eld to gravity, ). A direct evaluation of the induced parameters by computing the relevant Feynman loops gives L% = Lkinetic + g U L RH + h:c: + ZH |9( H |2 − m2H H † H − ZH = g 2 Nc log(2 =%2 ); (4)2 m2H = 2 − 10 = 2g4 Nc log(2 =%2 ); (4)2 )0 = 2g2 Nc 2 ( − %2 ) ; (4)2 1 g 2 Nc log(2 =%2 ) 6 (4)2 (B.5) (the parameter g is unrenormalized at this stage in fermion loop approximation). The induced low energy parameters, ZH and 10 , and )0 are determined in terms of , and we are interested in the % ≈ 0 limit of the theory. We emphasize that the eQective theory applies in either the spontaneously broken or unbroken phases. The broken phase is selected by demanding that m2H ¡ 0 for scales %2 2 , thus requiring that 2 (1−g2 Nc =82 ) ¡ 0; hence, g2 ¿ 82 =Nc =gc2 de 162 ; Nc log(2 =%2 ) m˜ 2H = m2H =ZH ; 1˜ = 10 =ZH2 = 322 ; Nc log(2 =%2 ) ) = )0 =ZH = 1=6 : (B.7) These are the physical renormalized coupling constants. We tune the low energy value of m˜ 2H to the desired value, and the remaining predictions of the model are contained in g, ˜ 1˜ (and )) as we will see below. The compositeness of the H boson is essentially contained in the result that g and 1 are singular as % → (while ) is constant and equal to its conformal value of 1=6). We will refer to these as the “compositeness conditions”. These results are easily recovered directly from the conventional diQerential renormalization group equations, supplemented with “compositeness conditions” as high energy boundary conditions. 384 C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 We utilize the approximate -functions which reIect only the presence of fermion loops: 162 9 g = Nc g3 ; 9 ln % (B.8) 162 9 1 = (−4Nc g4 + 4Nc g2 1) : 9 ln % (B.9) Solving the (B.10) If we now use the boundary condition, 1=g2 ()=0 we see that we recover the above result g˜ 2 =gc2 (%). The second RG equation may then be solved by hypothesizing an ansatz of the form 1 = cg2 . Substituting one 1 9 g= (4c − 4)Nc g3 9 ln % 2c (B.11) and demand that this must be consistent with the other RG equation. Thus one (B.12) and again 1()−1 = 0 leads to the above result of 1˜ = 1(%). To obtain the phenomenological results of the NJL model we examine the low energy Higgs potential from the action with % = m: 1˜ V (H ) = −m˜ 2H H † H + (H † H )2 − (g˜ U L 2 RH + h:c:) : (B.13) √ Let us assume that m˜ 2H ¡ 0 so the neutral Higgs <eld develops a VEV: Re(H 0 ) = (v + h)= 2. Therefore we (B.15) and so ˜ g˜2 = 4 m2h =m2 = 21= or mh = 2m (B.16) using Eqs. (B.7). This is the familiar NJL result, mh = 2m. We note that this result is subject to renormalizion when eQects of other interactions are included [383,540]. It also does not imply that the Higgs is loosely bound (see below)! We also obtain presently the Pagels–Stokar relation: 1 2 Nc v = m2 = g˜ 2 = m2 ln(2 =m2 ) : 2 162 (B.17) C.T. Hill, E.H. Simmons / Physics Reports 381 (2003) 235 – 402 385 It is also amusing to study the result ) = 1=6 in the diQerential renormalization group [540]. One (B.18) is a RG constant for all scales. More generally, as one descends toward the infrared, ) = 1=6 is an attractive infra-red Cxed point. Therefore, no matter what is the initial value for ) at the large scale , given enough RG running time ) will eventually reach 1=6 for small %. Of course, the RG running only occurs for scales % ¿ mH . We thus Y. Nambu, Axial vector current conservation in weak interactions, Phys. Rev. Lett. 4 (1960) 380–382. J. Goldstone, Field theories with ‘superconductor’ solutions, Nuovo Cimento 19 (1961) 154–164. M. Gell-Mann, M. Levy, The axial vector current in beta decay, Nuovo Cimento 16 (1960) 705. M. Gell-Mann, Phys. Rev. 92 (1953) 883. M. Gell-Mann, Nuovo Cimento 4 (Suppl. 2) (1956) 848. G. Zweig, An SU (3) model for strong interaction symmetry and its breaking, 2, CERN-TH-412. J.D. 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