Research Notes in Mathematics Subseries in Mathematical Physics Advisory Editors: R G Douglas, State University of New York at Stony Brook R Penrose, University of Oxford
D E Lerner & P D Sommers (Editors> University of Kansas & North Carolina State University at Raleigh
Complex manifold techniques in theoretical physics
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© D E Lerner and P D Sommers 1979 AMS Subject Classifications: 81XX, 14XX, 53XX AI! rights reserved. No part of this publication may be reproduced, stored in a retrieval system, or transmitted in any form or by any means, electronic, mechanical, photocopying, recording and/or otherwise without the prior written permission of the publishers. The paperback edition of this book may not be lent, resold, hired out or otherwise disposed of by way of trade in any form of binding or cover other than that in which it is published, without the prior consent of the publishers. Manufactured in Great Britain ISBN 0 273 08437 2
Preface
Over the past few years, a remarkable relationship between complex analysis and mathematical physics has emerged.
Through the use of Penrose's twister
theory, the solutions to some fundamental differential equations of mathematical physics can be converted into complex analytic objects on projective space.
The corresponding classification problems of complex geometry can,
in principle, be solved completely, with the result that one now has an explicit description of the set of all solutions to the (selfdual) YangMills, Einstein, and zero restmass equations. The articles in this volume are based on talks presented at a workshop on this subject held at Lawrence, Kansas from July 10 through July 15, 1978. The principal speakers at the workshop were D. Burns (Princeton), R. Hartshorne (Berkeley), R. Jackiw (M.I.T.), E. T. Newman (Pittsburgh), R. Penrose (Oxford), R. S. Ward (Oxford), and R. 0. Wells, Jr. (Rice).
In
addition to the articles by the principal speakers, these proceedings also contain several articles of a more technical nature contributed by the workshop participants. The workshop was supported by the National Science Foundation under the auspices of the Rocky Mountain Mathematics Consortium. organizers were D. Lerner and P. Sommers.
The workshop
Contents
Introduction
1
R. Jackiw Nonlinear equations in particle physics R.
s.
5
Ward
The selfdual YangMills and Einstein equations
12
R. Hartshorne Algebraic vector bundles on projective spaces, with applications to the YangMills equation
35
N. H. Christ Selfdual YangMills solutions
45
R. Penrose On the twistor descriptions of massless fields
55
R. 0. Wells, Jr Cohomology and the Penrose transform
92
L. P. Hughston Some new contour integral formulae
115
C. M. Patton Zero rest mass fields and the Bargmann complex structure
126
D. Burns Some background and examples in deformation theory
135
E. T. Newman Deformed twistor space and Hspace
154
K. P. Tod Remarks on asymptotically flat Hspaces
166
J. Isenberg and P. B. Yasskin Twistor description of nonselfdual YangMills fields
180
E. Witten Some comments on the recent twistor space constructions
207
J. Harnad, L. Vinet and S. Shnider Solutions to YangMills equations on ~ M invariant under subgroups of 0(4,2)
219
P. Green Integrality of the Coulomb charge in the line space formalism
231
N. H. Christ Unified weak and electromagnetic interactions, baryon number nonconservation and the AtiyahSinger theorem
234
Introduction
The differential equations of classical and quantum physics have provided a fertile common background for collaboration between mathematicians and physicists for many years.
In particular, the desire to understand linear
differential equations has led the mathematician to significant advances in the areas of functional analysis, differential and complex geometry, and the theory of Lie groups.
Many of these same developments have given the
physicist a deeper understanding both of the successes of the quantum theory and of the problems remaining to be solved. In the case of nonlinear equations, however, the results have been less spectacular.
This is particularly unfortunate from the physicists' stand
point, for these are the equations related to the description of nontrivial interactions.
A case in point at the classical level is that of general
relativity, in which the gravitational field is coupled to itself through the nonlinear terms in Einstein's equations:
On
the one hand, there are
the shorttime existence and uniqueness theorems for the Cauchy problem; on the other, a finite number of families of exact solutions, many of which appear to be devoid of physical significance.
There is no known method of
constructing the general local solution from free data.
The only information
available as to the qualitative nature of global solutions suggests that, in the generic case, singularities develop and the solutions do not persist for all time.
It goes without saying that our poor understanding of the
classical fields does not make the quantum mechanical problem any easier.
1
Einstein's equations illustrate several typical problems which arise whenever one is confronted with a system of nonlinear partial differential equations: (a)
What is the nature of a generic local solution?
How does one
construct solutions explicitly? (b)
What can be said concerning the set of all solutions (possibly sub
ject to some suitable boundary conditions)?
Can this set even be defined in
a rigorous way, and, if so, how is it parametrized? These are extremely difficult questions, and it is a remarkable discovery that the mathematical techniques routinely employed by algebraic geometers and complex analysts can, in some cases, be brought to bear on these problems with stunning effect.
In several important instances, the solutions to a
given system of field equations can be represented entirely in terms of complex analytic objects.
Very roughly speaking, the field equations can
be converted into the CauchyRiemann equations by making suitable changes in the geometric background space. There are, at the moment, three concrete examples of this phenomenon: Massless free fields:
The positive frequency fields of a definite heli
city have been shown by Penrose to be isomorphic to certain sheaf cohomology groups over a fixed open subset of projective threespace, P 3 (c).
Any real
analytic field can be similarly represented, and nonanalytic fields can be represented in terms of relative cohomology groups and hyperfunctions, as shown by Wells. YangMills fields:
Ward has shown that any realanalytic selfdual Yang
Mills field defined in an QRen region of Minkowski space determines a holomorphic vector bundle on an open 8ubset of P 3 (c).
The original field can be
reconstructed from the transition functions of the bundle.
2
The (Euclidean)
SU(2) YangMills fields describing pseudoparticles or instantons are classified by stable rank2 vector bundles on P 3 (c) satisfying a certain reality condition.
The ninstanton solutions are parametrized by the points
of a real algebraic variety of dimension Sn3
(Atiyah, Hitchin, Singer, and
ward). Halfflat solutions to Einstein's equations:
These are (necessarily
complex) fourmanifolds with Ricciflat holomorphic metrics whose curvature tensors are either selfdual or antiselfdual.
Penrose has shown that any
local solution sufficiently close to flat space can be obtained by deforming the complex structure of a neighborhood of a line in P 3 (c) in an appropriate way.
Starting from an altogether different point of view, Newman has shown
how to associate a halfflat spacetime with any realanalytic asymptotically flat vacuum solution sufficiently close to Minkowski space. context, the halfflat spacetime is called an Hspace.
In this
Newman's construction
also relies on deformation theory, and the two approaches are closely related While these examples differ significantly in their details, in each case the set of solutions is in onetoone correspondence with a well defined class of complex analytic objects.
The differential equations themselves
have disappeared, having been in some sense coded into pure complex structure; and the problem of enumerating and classifying the solutions has been converted into a problem in complex geometry (e.g., classifying a subset of the stable rank2 vector bundles). Since these results are quite recent, only the barest outlines of what may eventually become a systematic and powerful theory are understood at the moment.
Even so, the subject appears quite promising for the following
reasons:
3
(1)
The physical problems which have been successfully handled so far
are nontrivial and of considerable current interest.
Moreover, the
techniques involved in the constructions were developed, not for their mathematical generality, but to deal specifically with problems of physical significance.
It is reasonable to expect that further developments in the theory
will have direct physical applications.
It is even possible that some of
the long standing problems of quantum field theory may be solved by a reformulation involving complex manifolds. (2)
From the mathematician's point of view, this is a largely unexplored
area with great potential.
The examples cited above indicate the existence
of a class of differential equations which can be effectively treated by complex manifold techniques.
It is perhaps worth emphasizing three of the
common features of these examples: (a)
The methods are constructivegiven the appropriate complex
analytic object, one can in principle, and frequently in practice, proceed to construct a solution in closed form. (b)
The general local solution is obtained in each case.
(c)
Certain types of boundary conditions, such as finite action or
positive frequency, can be imposed on the global solutions by properly restricting the class of complex analytic objects. It is clearly an important matter to determine the full generality of these methods. D E LERNER
Department of Mathematics University of Kansas Lawrence, Kansas 66045
P SOMMERS Department of Mathematics North Carolina State University Raleigh. North Carolina 27650
R Jackiw Nonlinear equations in particle physics
In the last few years, a new approximation procedure has been developed for analyzing quantum field theory.
This method has allowed physicists who are
examining various field theoretical models to get results which had not been anticipated from earlier, perturbative investigations.
The new approach
begins with a study of the relevant field equations, viewed as partial differential equations for ordinary functions (i.e. for classical fields) rather than as Heisenberg field equations for quantum operator fields.
The field
equations are analyzed and solved by techniques of mathematical physics, and then the quanta! information is extracted from these classical solutions. In other words, quantum field theorists have learned what quantum mechanicians of the 1920's knew already:
semiclassical methods can give an
approximate quanta! description of a physical system, provided the classical description of that system is completely known; this is just as true for modern quantum field theory as it is for particle quantum mechanics. In order to make progress with the new program, physicists have been solving partial, nonlinear differential equations which arise in this contex.t.
It happens that these equations are also mathematically signifi
cant; hence we are witnessing a fortunate conjunction of interest between physicists and mathematicians, and that is why a physicist is here addressing an audience of mathematicians. Several reviews of the subject are available; consequently this written record of •Y lecture will be brief and the interested
r~der
is referred to
the following extensive summaries.
5
S. Coleman, "Classical LUDlps and their QuantUDl Descendants" in New Phenomena in SubNuclear Physics, edited by A. Zichichi, Plenum Press, New York (1977).
s. ~·
Coleman,"Uses of Instantons 11 in edited by A. Zichichi, PlenUDl Press, New York (in press).
R. Jackiw,
11
Quantum Meaning of Classical Field Theory"
Reviews of Modern Physics 49, 681 (1977). R. Jackiw 1 C. Nohl and C. Rebbi,
11
Classical and semiClassical Solutions
of the YangMills Theory 11 in Particles and Fields, edited by D. Boal and A. Kamal, Plenum Press, New York (1978). A very short account is by R. Jackiw and C. Rebbi 1
11
Topological
St.:~litons
and Instantons 11
Comments on Nuclear and Particle Physics (in press). Of course the partial differential equations possess a vast variety of solutions, and only certain special ones have been used thus far by physicists to probe the quantum theory. and the
11
These are the "soliton11 solutions
pseudoparticle11 or "instanton11 solutions.
Here I shall describe
them and explain their significance for physical theory. The solitona are real solutions with finite energy and localization in space.
Also they are stable against small deformations.
(This definition
does not coincide with usage in the engineering and applied mathematics literature.)
Their behavior at large spatial distances frequently is topo
logically nontrivial; in that case, no finiteenergy evolution can connect a solution with the soliton's asyaptotes to a solution with the vacuum's asyaptotes. 11
There is an infiniteenergy barrier separating the two.
vacuUDl11 is the zeroenergy solution of the equations.)
6
(The
For the quantum
theory, these classical soliton solutions are interpreted as evidence that there exist unexpected particle states, the socalled "quantum solitons," which are neither the "elementary" particles of the model, nor are they bound states of a finite number of elementary particles.
Rather quantum
solitons are bound states of an infinite number of elementary particles; they represent collective, coherent excitations.
Quantum soliton states
arise in various field theoretical models, but at present it is not known what phenomenological significance should be assigned to them.
The theory of
soliton solutions is discussed in the first and third reviews cited above. Instantons (pseudoparticles) are real solutions of field equations that are continued to imaginary time (t
~iT).
It has been established that
such solutions can be used to give an approximate, semiclassical description of quantum mechanical tunnelling.
Numerical computation of the
tunnelling amplitude proceeds according to the following sequence of steps. Firstly the quantum field theory is formualted in· terms of npoint functions (vacuum expectation values of the timeordered products of n quantum fields). These npoint functions can be obtained from a generating functional for which there is a functionel integral representation.
Next the functional
integral is continued from Minkowski space to Euclidean space.
Finally, the
Euclidean space tonctional integral is evaluated approximately by expanding about the Euclidean (imaginarytime) solutions. Instanton solutions and calculations of tunnelling have been principally performed for YangMills theorythe field theory which is currently believed to be appropriate for a correct description of fundamental processes in Nature.
A phenomenologically important result has thus been achieved:
tunnelling reduces the degree of symmetry so that physical predictions of the theory possess less symmetry than had been anticipated.
Precisely such a
7
·reduction of symmetry is required to obtain agreement with experimental data. The second and fourth reviews mentioned above are devoted to this topic. YangM1lls instanton solutions can be labelled by their Pontryagin index. If AP is an SU(2) gauge potential in Euclidean space (a = 1, 2, 3, internal a symmetry indices; p = 1, 2, 3, 4, Euclidean space indices) and Fpv is the a 1 corresponding gauge field, then the Pontryagin index is q = d 4x *Fpv F , 32u 2 a apv where *Fpv is the dual gauge field, related .to the gauge field by the antia *FPV _ ! pvaaF symmetric tensor, a  2 £ aaa· For sufficiently wellbehaved fields,
f
q is an integer, and for the known instanton solutions it takes on arbitrary integral values.
Moreover, these solutions are selfdual <*Fpv a
Both physicists and mathematicians have classified all selfdual instantons and have shown that when q = n, the solution depends on Blnl3 gauge invariant parameters (for the SU(2) group; analogous formulas hold for arbitrary groups). follows:
The physical interpretation of these parameters is as
the q = n instanton can be thought to be a nonlinear super
position of lnl qa±l instantons. specify it:
Each instanton requires 8 parameters to
4 determine the location in Euclidean space, 1 parameter sets
the instanton's size, while the remaining 3 give its orientation in group space.
The total Bini is decreased by 3, which represents an overall gauge
rotation of the entire ensemble.
Physicists have given explicitly solutions
with 5lnl+4 parameters; a constructive procedure for obtaining the most general Blnl3 parameter solution has been found by mathematicians, and is described in this conference by Ward. While these results have shed much light on the mathematical structure of YangMills theory, their physical import has not as yet been fully established.
Only the lnl = 1 solution has played a role in practical, physical
calculations; those with lnl > 1 appear to give only negligibly small 8
corrections.
Thus one open question concerns the proper role of instantons
with. lnl > 1 in physical theory.
Other interesting questions which remain as
yet unanswered include the following.
What are the physically relevant
regularity conditions at infinity which render the action (:
41 fd 4x
F~vF a
a~v
) finite and the Pontryagin index q integervalued?
known instantons have finite action and integervalued q.)
(The
What distin
guishes solutions of YangMills equations that are not selfdual and do they have significance for physical theory?
(Mathematicians have recently shown
that stable solutions are necessarily selfdual.)
What are the properties of
the Blnl3 parameter space of solutions and does there exist a group of symmetry transformations acting on these parameters? answer is known.
=1
the
The 15 parameter conformal group 0(5,1)a symmetry group
of the YangMills theoryacts on the 5 parameter lnl follows:
(For lnl
=1
solution as
the 10 parameter 0(5) subgroup leaves the solution invariant; the
remaining 5 parameter family of transformations medifies the parameters of the solution.)
Some progress on this question is here reported by Hartshorne.
Finally, impressed by the marvelous structure of Euclidean YangMills theory, some of us have returned to the Minkowski space theory and looked for classical solutions there.
It is known that no soliton
(nondissipative.
particlelike) solutions exist. but other kinds of solutions have been found. It has proved useful to classify them according to their transformation properties under the Minkowskispace conformal group, 0(4,2). and its subgroups.
Especially interesting are real solutions invariant under the
0(4) x 0(2) subgroup. as well as those invariant under the 0(4) subgroup. Complex solutions. invariant under other subgroups, have also been exhibited and are described at this meeting by Shnider.
It has been suggested that
the real solutions may be used to determine the semiclassical spectrum of
9
various conformal generators (see the fourth reference}.
The role of complex
solutions in the physical theory is not known. Another body of calculation that has interested physicists and mathematicians concerns the coupling of an external, linear system to a soliton or instanton; for example one studies the Dirac or KleinGordon equations in the field of an instanton or soliton.
In this way one is led to examine
linear, elliptic differential operators which frequently possess vanishing eigenvalues.
(Small deformations of solitons or instantons satisfy equations
of this type as well.)
The zeroeigenvalue modes have profound significance
for the physical theory; their role is explained in the third and fourth cited reviews.
Close study of these modes has also put physicists in contact
with mathematicians since the principal mathematical result relevant here is the AtiyahSinger Index Theorem or equivalently, as physicists came to know it, the anomaly of the axial vector current. The semiclassical method, based on classical solutions, has provided new understanding of quantum field theory.
However, we must remember that a
complete reconstruction of the quantum theory from classical·solutions is not in general possible.
Certainly the restriction to finite energy or finite
action, which has characterized the solutions emphasized by all of us thus far, is only appropriate to approximate quantum mechanical calculations; classical solutions with infinite energy and/or action are also needed for understanding the full quanta! system. integral
r~presentation
Moreover, recall that a functional
of a quantum theory involves a functional integration
over all classical configurations, not just those satisfying equations of motion.
Hence, in general, we shall have to go beyond classical solutions
and examine classical field configurations to obtain further insights into the quantum theory.
10
~knowledgments
This summary of my lecture was prepared at the Aspen Center for Physics and is based on notes taken by Professor B. DeFacio, whose help is gratefully acknowledged.
The research is supported in part through funds provided by
the U.S. Department of Energy (DOE) under contract EY76C023069.
R JACKIW Center for Theoretical Physics Laboratory for Nuclear Science and Department of Physics Massachusetts Institute of Technology Cambridge, Massachusetts 02139
11
R SWard The selfdual YangMills and Einstein equations
§1.
INTRODUCTION
As was pointed out by Jackiw in his lecture, it is often useful to study the classical (i.e. nonquantum) versions of quantum field theories. classical system is YangMills theory.
One such
Another classical theory (which does
not, as yet, have a satisfactory quantum version) is Einstein's theory of gravity.
The YangMills equations and Einstein's equations are hyperbolic
systems of nonlinear equations in spacetime.
In this lecture we shall be
working with their analytic extension to complexified spacetime and, in particular, to positive definite 4space (so that they become elliptic rather than hyperbolic). The YangMills equations and Einstein equations, being nonlinear, are by no means easy to solve.
This is where complex manifold techniques come in.
When one transforms the problem from spacetime to a certain complex manifold called twistor space, one discovers that the selfdual solutions of the YangMills and Einstein equations can all be obtained without having to solve any differential equations at all.
This transformation (the socalled Penrose
transform) is the subject of the present lecture. Twistor theory was invented and developed by R. Penrose; in particular, it was he
wh~
produced the crucial theorems 2.3 and 5.1 discussed below.
The
application described in §§3 and 4 of twistor formalism to YangMills theory is due mainly to M.F. Atiyah and myself. To begin with, we need to establish some notation.
Let R denote the real
numbers, C the complex numbers and C* the multiplicative group of nonzero
12
complex numbers.
Complexified
Minko~ski
spacetime CM is C4 equipped with a
flat, nondegenerate, holomorphic metric g. [In other words, there exist complex coordinates x
a
(a
= 0,1,2,3)
on CM such that g is given by the complex
line element
clearly CM can also be thought of as the complexification of Euclidean 4space E4 •
LetT denote the holomorphic tangent bundle of CM and An the bundle of
holomorphic nforms.
The tangent space at x
E
CM is denoted T , and so forth. X
Choose a constant volume element (4form) on CM; this, together with the metric, defines a duality operator
*:
2
A
2
+A,
with ** = identity.
Then A2 may be decomposed into orthogonal subspaces:
2+
where A  is the eigenspace of * corresponding to the eigenvalue ±1.
The
2+ 2sections of A are called selfdual 2forms, while those of A are anti
selfdual. Now let us introduce spinors.
There are two basic spinbundles, denoted
SandS'; their duals are denoted S* and S'* respectively.
Each of them is
a 2complexdimensional vector bundle over CM with structure group SL(2,C); this means that each fibre Sx, x E CM, is equipped with a symplectic form E. Spinors are tied to the background geometry in the following way.
First,
one chooses an isomorphism T
~sas•.
(1.1)
13
The dual version of this is (1.2)
We require that the metric g and the symplectic forms £ and £ 1 on S and S' be related according to g
(1.3)
£ 8 £'.
where the equality sign is to be understood in terms of the isomorphism It follows from this a vector v in T
(1. 2).
X
iff it is decomposable inS
o~
g(v,v) where A
€
s, X
~ €
v
=
X
8 S'. X
is null (i.e. has zero length)
In other words,
A 8 ~.
(1.4)
s•. X
Let S(n) denote the symmetrized tensor product of n copies of S, and similarly for S'(n), S*(n) and S'*(n).
From (1.2) and the
twodimensionalit~
of spinspace i t follows that A2 : S*(2) • S'*(2). This is
t~
2
decomposition of A into selfdual and antiselfdual parts 
but we still have to choose which part is which. !i. 2+
=
A2
= S*(2).
S'*(2),
We make the choice (1.5) (1.6)
Each spinbundle is equipped with a connection V :
s
+
s
1 8 A ,
etc. In CM all these connections are flat.
(1. 7)
Complex projective mspace is denoted P •
Let L(l) be the hyperplane
m
section bundle over P
m
and let L(n) be the line bundle [L(l)]n.
Let O(n)
denote the sheaf of germs of holomorphic sections of L(n). (So when written in terms of homogeneous coordinates, sections of O(n) correspond to holomorphic functions homogeneous of degree n.)
Since L(O) is the trivial
bundle, 0(0) is the same as 0, the sheaf of germs of holomorphic functions on P • m
One final bit of notation: if E is a holomorphic vector bundle,
then f(E) denotes the group of holomorphic sections of E.
§2.
TWISTOR GEOMETRY AND MASSLESS FREE FIELDS
We begin with a description of the geometrical correspondence between complexified spacetime and twistor space; for more details, see [9]. An aplane is defined to be a complex 2plane Z in CM with the following two properties. let
~
Let u and v be any two vector fields on Z, tangent to Z;
and v be the corresponding 1forms.
Then
(i)
Z is totally null \i.e. isotropic): g(u,v) = 0;
(ii)
Z is selfdual:
~ A
v is a selfdual 2form.
Similarly, a aplane is a totally null antiself dual 2plane in CM. what follows we shall be using aplanes rather than aplanes.
In
But every
statement has an "aversion" in which "selfdual" and "antiselfdual" are interchanged. Proposition 2.1
The space of aplanes in CM is the quotient of S' 0
X
(S*{0}) 0
by the natural multiplicative action of C*.

Proof
S' etc. means spinors at the origin: we are using the flat connection 0
on CM to refer everything to the origin.
A point x e
~~
is represented by
15
its position vector with respect to the origin: X €
S M S' =Hom (S* S'). o o o' o
T
0
So given (w,n)
w
E
S'0 x (S*0  {0}) • the equation
(2.1)
x.n
For fixed (w,n), the solution space of (2.1) in x is a Bplane
makes sense.
(this follows from equations (1.4) and (1.6)).
Every Bplane arises in this
'V'V
way; (w,n) and (w,n) determine the same Bplane iff w =A~
'V
•
Now (S' x S*)/C* 0
0
TI = ATI
for some A
E
0
C* •
~ c 4/C* = P 3 , so Proposition 2.1 tells us that the space of
Bplanes in CM is P 3  I, where I is the line in P 3 given by TI = 0.
Here
(and in what follows) "line" means "complex projective line"; so intrinsically I
~
P1 •
We say that P3  I is the projective
~istor
space corresponding to CM;
each Bplane Z corresponds to a projective ~stor
ZE
P 3  I.
[In the
literature these are usually called dual projective twisters, while the Next we consider what spaae
aplanes correspond to projective twisters.]
time points correspond to in twister space.
A point x
E
CM is characterized
by the set of all Bplanes through it: from equation (2.1) it follows that this set is the line
X
in p 3  I. as
16
{(w,n)ln f 0,
w
x.n}
So we can summarize the spacetime
++
twister space correspondence
point x
€
CM
line x in P3  I,
8plane Z in CM
point Z
++
P 3  I.
€
The space of all lines in P3 is a compact 4dimensional complex manifold, in fact a quadric Q in P 5 •
So CM (thought of as the space of lines in
p 3  I) is an open submanifold of Q; and Q is a compactification of CM. The crucial feature of the above correspondence is the way that it relates to the
confo~al
(i.e. nullcone) structure of CM: the following result is
easily verified. Two points x, y
Proposition 2.2
corresponding lines
CM lie on a common null geodesic iff the
€
x and y in P3
 I intersect.
In this lecture we shall be working with two variants of the CM
++
P3  I
correspondence, namely the local version and the global positive definite version.
In the local version we use an "open ball" X in CM.
By this is
meant that X is an open submanifold of CM, such that the intersection of X with any linear subspace of CM is topologically and ho?Jmorphically trivial. The corresponding region X of twistor space is defined to be the space of 8planes which inte·.:sect X. point x
€
X
8plane Z in X
So, in other words, we have the correspondence
++
line ~ in
X,
++
" point Z
"X.
€
" has the topology S2 x R4 (as opposed to X, which is of course The space X
topologically trivial). For the positive definite version we restrict our attention to a Euclidean "slice" E4 of CM.
The structure group on the spinbundles over E4 is reduced
from SL(2,C) to SU(2). complex structure o:S
This means that each spinspaceS is equipped with a ~
S, such that
17
e:(.
,o.)
is a positive definite Hermitian form. Now o induces an action on T =Sa S', the fixed points of which are precisely the real vectors, i.e. the real points E4 in CM.
It also gives rise
to an antiholomorphic involution, again denoted o, on P3 on P3  I.
This map o: P3 + P3 has no fixed points.
fixed lines: let us call these "real" lines.
the PeaZ points x
£
= (S'
x
S*)/C* and
However, it does have
Thus
E4 coPPespond to the PeaZ Zines x in P3  I.
In addition, the line I is a real line, so the real lines in P3 correspond to the points of the onepoint compactification
s4
of E4 •
Clearly no two
distinct real lines can intersect each other (if they did, then the corresponding points on the metric on
s4
s4
would be nullseparated; but this is impossible, since
is positive definite).
So PJ is a bundle oVeP
s4 ,
with the
real lines as fibres. It should be remarked that real Minkowski spacetime (signature +) can be obtained by imposing a different "real structure" on CM and P3 •
The spin
bundles S and S' remain SL(2,C)bundles, but become complex conjugates of each other.
The upshot of this is that every selfdual real 2form in
Minkowski spacetime is also antiselfdual, and hence vanishes. Finally in this section we discuss massless free fields. and (1.7) imply that the derivative operator sections of S*(n)
a
S* aS'*·
v maps
Equations (1.2)
sections of S*(n) to
Now S*(n) a S* decomposes into its irreducible
parts as follows: S*(n) a S* Let
T
S*(n+l)
S*(n1).
be the projection from S*(n) a S* a S'* onto S*(n+l) a S'* and
projection from S*(n) aS* 18
$
a
S'* onto S*(n1)
a
S'*·
Then
~
the
f(S*(n))
~
is called the twistor
operator~
f(S*~)) ~
r,.o'iJ
(2.2)
f(S*(n+l) 9 S'*) and
(2.3)
f(S*(n1) 9 S'*)
is called the Diraa operator.
Both these operators have "primed" versions;
for example, there is a primed Dirac operator r,.'oV
f(S'*(n))
~
~
f(S'*(n))
f(S'*(n1) 9 S*).
A massless free field is a symmetric spinor field which is annihilated by the Dirac operator.
More precisely, let X be an open ball in CM and let n
Then the space
be any integer.
~
n
(X) of heliaity in massless free fields in
X is
(i)
if n > 0, the kernel of r,.'oV acting on sections of S'*(n) restricted to X;
(ii)
if n < 0, the kernel of r,.oV acting on sections of S*(n) restricted to X;
(iii) if n
= 0,
the kernel of 0 (the covariant Laplacian/d'Alembertian)
acting on scalar fields in X. Theorem 2.3
The vector spaces
~
n
"' (X) and H1 (X,O(n2)) are naturally iso
morphic. Remark
For a proof, see Penrose's article in these proceedings.
The
theorem developed out of the wellknown Kirahoff integral formula for massless free fields.
This integral formula (in real Minkowski spacetime)
expresses the value of the massless field at the point x as an integral of the initial data over some crosssection of the past null (i.e. characteristic
19
cone of x.
In general, for a hyperbolic equation, no such formula exists,
because there is usually a contribution from inside the past characteristic cone.
The existence of the Kirchoff formula for massless free fields re
fleets the fact that they satisfy Huygens' principle (in fact, this is the usual definition of Huygens' principle). Among other equations satisfying Huygens' principle are the selfdual YangMills equations and (in some sense)
th~
selfdual Einstein equations.
As we shall see, both these equations can be treated very satisfactorily within the twister formalism.
In fact twistor theory seems to have a pre
ference for equations satisfying Huygens' principle: this is hardly surprising, in view of the way in which the spacetime
++
twistor space correspondence is
tied up with nullcone (i.e. conformal) geometry of spacetime.
A corollary
of this is that equations which do not satisfy Huygens' principle (unfortunately these include the full nonselfdual YangMills and Einstein equations) are likely to be much more difficult to handle.
For a discussion
of the full YangMills case, see the article by Isenberg and Yasskin in these proceedings.
§3.
THE SELFDUAL YANGMILLS EQUATIONS
By way of motivation, let us consider a special case of Theorem 2.3, namely the case n
= 2:
~ 2 (X) ~
1 ,..
H (X,O).
Now the isomorphism (1.5) implies that
~ 2 (X)
is precisely the space of closed
selfdual 2forms in X  these are called selfdual exact sequence of sheaves 0 + Z c.. 0
20
e~ 0* + 0
M~ell
fields.
The short
(where Z is the sheaf of integers and 0* the sheaf of nowherezero bolomorphic functions) leads to a long exact sequence of cohomology groups, part of which is
0
+
H1" (X,O)
+
v 2" H1" (X,O*)~ H (X,Z)
+
0.
1 .... .... The group H (X,O*) is the group of hotomorphic tine bundtes over X and the 2 .... map y maps a bundle to its Chern class in H (X,Z). So we arrive at the
following result. Proposition 3.1
Selfdual Maxwell fields in X correspond in a 11 fashion .... to bolomorphic line bundles over X with vanishing Chern class. A Maxwell field, being a closed 2form in X, is usually interpreted
geometrically as the curvature of a connection on a line bundle over X.
We
can generalize Maxwell theory (and Proposition 3.1) by simply going from line bundles to vector bundles.
The resulting theory is called gauge theory
or YangMills theory; its basic ingredients will now be summarized. Let X be an open ball in CM bundle over X.
and let E be an rdimensional complex vector
Let A be a linear connection on E: in the language of physics
A is a gauge potentiat.
If we choose a basis for E (i.e. r linearly
independent sections of E), then A may be thought of as an r x r matrix of 1forms on X.
A section of E is represented as a column rvector
~
of
functions on X, and its covariant derivative is the vectorvalued 1form Dljl : = Vlji
+
A~.
(3.4)
A transformation to a different basis for E is known as a gauge transformation. Here the gauge group (or structure group) is GL(r,C).
One can specialize to
some other gauge group G c GL(r,C) by requiring that A be a Gconnection, i.e. that A be a 1form with values in the Lie algebra of G. 21
The gauge
fie~
F is the curvature of A, represented as a matrix of
2forms: F = dA + [A,A] •
(3.2)
From (3.2) it follows that the curvature F automatically satisfies the Bianchi identities dF + [A,F]
o.
(3.3)
The YangMi Us equations look very similar:
o.
(3.4)
Clearly if F is selfdual, i.e. if
then the YangMills equations (3.4) follow automatically from (3.3).
So the
selfdual gauge fields form a subclass of the space of solutions of the YangMills equations. If Z is a
Bplane in X, we can restrict the bundle, connection and
curvature to Z: E t+ E(Z), Proposition 3.2
At+ Az•
F
I+
Fz.
The gauge field F is selfdual iff AZ is trivial for every
Bplane z. Proof
The connection AZ on E(Z) is trivial iff its curvature FZ vanishes.
But the Bplane Z is antiselfdual, so
22
where F is the antiselfdual part of F. a spinor
~
e 5*(2).
Now F
corresponds, by (1.6), to
Using the representation of a twistor Z as a pair of
spinors (w,n), one can show that (F)z = 0 for all z 4===9 ~(n,n)
for all
1T
e S
=0
~
~
~
F is selfdual.
0
There is a 11 correspondence between
Theorem 3.3 (a)
0
connections on the rdimensional bundle E over X, with selfdual curvature; and
{b)
holomorphic rdimensional vector bundles
Eover X,
satisfying the
condition
(Here
E(i)
is holomorphically trivial, for all points x e X.
E(;)
means "the bundle
Sketch of proof
Erestricted
to the line
i
in
(3.5)
X".)
Let us describe geometrically what the correspondence is.
To begin with, suppose we are given the bundle E with a selfdual connection. Define K(Z) to be the space of covariantly constant sections of E(Z), i.e. K{Z) = {ljl
E
r(E(Z)) lv...JDljl = 0 for all vectors v tangent to Z}.
" by defining its fibre EA " over an arbitrary point Z ,... Now we define E
z
E
" X:
A
EA
K(Z).
z
Since K{Z) that
Eis
=
Cr by Proposition 3.2, it is not difficult to convince oneself
an rdimensional holomorphic vector bundle over
X.
To establish
that EA satisfies (3.5), it is sufficient to find a biholomorphic isomorphism
;:E(~)
+ ;
x E • X
Such a map
~
can.be defined as follows.
Pick
Z E i.
A
23
point ~"
" E" corresponds by definition
€
z
section~
to a constant
of E(Z).
Now define l; by "
~
l;
where
~(x)
I+
(Z,~(x))
€
" X
means the section
Conversely, given
x Ex' ~
evaluated at x.
E satisfying
(3.5), define E by
the space of holomorphic sections of E(~).
Since E(~) is trivial and xis a ,.. ,..
Riemann sphere, Liouville's theorem implies that f(E(x)) rdimensional vector bundle over X. connection on E.
~
~(x) [Z)
E:
E,.,
z
So E is an
We do this by specifying the covariantly constant sections
maps a point x
section of E(~).
r
C •
It remains to define a selfdual
Suppose~
of E(Z), where Z is an arbitrary Splane. Thus
~
€
Z into Ex
is a section of E(Z).
f(E(~)): in other words, ~(x) is a
Now evaluate this section at
Z € i:
this gives
Then we say that ~ is a covariantly constant section of E(Z)
iff ~(x)[Z] has the same value for all x
€
Z.
Remarks (a)
Notice that for r
= 1,
Theorem 3.3 reduces to the Maxwell case
(Proposition 3.1). (b)
The difference between the spacetime and the twistor space pictures is that in the spacetime X, the bundle E is trivial and all the information is contained in the aonneation A; in the twistor space X, however, there is no connection: all the information is contained in the (nontrivial) bundle E.
(c)
Theorem 3.3 is a "local" theorem (in the sense of X being an open ball in CM); it has global versions, one of which we shall meet in §4.
(d)
It also has versions for gauge groups other than GL(r,C).
For example,
we could reduce the group to SL(r,C), as follows. There is a 11 correspondence between
Theorem 3.4 (1)
selfdual SL(r,C)connections on E;
(2)
holomorphic bundles dat
Eis
Esatisfying
and
(3.5), as well as the condition
trivial.
~
~
(det E is the determinant line bundle of E: its patching function is the determinant of the patching matrix defining the vector bundle
§4.
E.)
VECTOR BUNDLES ON P3 AND INSTANTONS
Instantons (or pseudoparticles) are solutions of the YangMills Equations 4
(3.4) on 5 , with gauge group SU(2) [see Professor Jackiw's lecture].
More
general gauge groups are also of interest (e.g. SU(n), Sp(n), etc.), but for simplicity we shall stick to SU(2). most interest in
connec~ion
The solutions which have been of
with physical applications are the selfdual and
antiselfdual solutions (in fact, it is not yet known whether there are any solutions other than these).
So what we want is a theorem (analogous to 4
Theorem 3.3) which characterizes selfdual SU(2)connections over 5 • Recall from §2 that the antiholomorphic involution o acts on P3 and that 4
the "real lines" of P3 correspond to the points of 5 • Theorem 4.1 (1)
There is a 11 correspondence between
2dimensional vector bundles E over
s4
with selfdual SU(2)connections;
and (2)
~
2dimensional holomorphic vector bundles E over P3 satisfying the conditions
25
(i)
E(i)
(ii)
c 1 (E) (the first Chern class of E) vanishes;
is trivial for all real lines x;
(iii) a lifts to an antilinear map ~A
A
a:E + E,
with
~
a
= 1.
Remarks (a)
For a proof, see (2].
(b)
c1
(E)
= 0~
Eis
det
trivial.
In other words (cf. Theorem 3.4),
condition (ii) reduces the gauge group from GL(2,C) to SL(2,C).
Then
condition (iii) achieves the further reduction from SL(2,C) to SU(2). (c)
The only other topological invariant of E is c 2(E): this is called the
instanton number. (d)
E is
necessarily algebraic, so all instantons are rational.
In his lecture, Hartshorne describes two ways of constructing algebraic vector bundles over P3 •
The first is to associate vector bundles with curves
in P3 : some of the features of this procedure are discussed below.
The
second method (due to Horrocks and Barth) will not be pursued further here [see Christ's lecture in these proceedings and reference 1]. The idea of the first method is as follows. vector bundle over P3 •
E8
Then there exists a positive integer q such that
L(q) has a section which vanishes along a curve
therefore,
E8
line bundle.
A
Let E be a 2dimensional
r
in P3 •
On P 3 
L(q) is an extension of the trivial line bundle by another After tensoring with L(q) we arrive at the result that
restricted to P3 
r,
is an extension of L(q) by L(q).
correspond to elements of the cohomology group H1 (P 3
E,
Such extensions
r, 0(2q)) and hence
to certain helicity (1q) massless free fields on spacetime.
By starting
with the massless free fields and working backwards, one can therefore
26
r,
construct all selfdual instantons.
This assertion is restated in the
following pair of theorems. Let q be a positive integer and let X be an open ball in CM.
Theorem 4.2
Then there exists a nonlinear functional Aq from
~ 2 _ 2 q(X)
 N to the space
of selfdual SL(2,C)connections on the 2dimensional vector bundle E over X. ~ 2 _ 2 q(X)
Here
is the vector space of helicity (1q) massless free fields on
X and N is a certain subset of this space (see remarks (c) and (d) below). Theorem 4.3 way.
All selfdual
instanton connections A can be obtained in this
In other words: given A, there exist an integer q, an open ball X in
CM and a massless field given by Aq
~ ~ ~ 2 _ 2 q(X)
4 such that A (restricted to X n S ) is
[~].
Remarks The functionals A are described explicitly in [3,4].
(a)
q
This, together
with Theorem 4.1 and the algebraicgeometrical results, provides a proof of Theorems 4.2 and 4.3. (b)
Theorem 4.2 is stated in "local" form to emphasize the fact that the
Aq are essentially local in nature.
(A global version would be somewhat
trickier to state, if only because there are no nontrivial massless fields on
s 4 .)
(c)
As an example, let us consider the case q
[2].
= 1.
For more details, see
The metric and volume 4form on CM define an (essentially unique)
linear map
Identifying A2+ with the Lie algebra of SL(2,C), we define the functional
A1 by 27
where~£
~is
t (X) (i.e. 0
a scalar field on X, satisfying
in order for A1 to be welldefined, we need
~
0~ =
0).
Clearly
to be nowhere zero on X.
So
the subset N that has to be deleted from t (X) consists of those fields which 0
are somewhePe zero on X.
Taking
0~
s4
= 0 based on points of
this case, the curve (d)
r
~
to be a sum of elementary solutions of
gives the socalled 't Hooft instantons.
In
is a disjoint collection of real lines in P3 •
For q > 1 the picture is similar, but somewhat less straightforward
The subset N is again determined by a "somewhere zero" condition.
r
of the curves
The genus
increases with q [7] so the instanton solutions become more
difficult to construct explicitly. (e)
Let M(k) be the space of selfdual instantons with second Chern class
k > 0.
Then M(k) is a real algebraic variety of dimension Bk3.
Let M(k,q)
be the subspace of M(k) consisting of those instantons which can be derived from the functional Aq.
For each fixed value of k, there exists a value qk
of q such that [7]
But qk
+ ~
as k
instantons.
in other words, no finite value of q gives aU
The following table illustrates these comments. 1
2
3
4
5
>5
dim M(k,l)
5
13
19
24
29
5k + 4
dim M(k,2)
5
13
21
29
36
4k + 16
dim M(k)
5
13
21
29
37
8k  3
k
28
+ ~;
Much work remains to be done on investigating the spaces M(k); most recent investigations have concentrated on the Horrocks/Barth approach mentioned above.
§5.
SELFDUAL SOLUTIONS OF EINSTEIN'S EQUATIONS
Let X be a 4dimensional manifold equipped with a nondegenerate metric g. may either be real (in which case g is g is to be complex holomorphic).
X
to be real) or complex (in which case
In either case, we say that g is selfdual
if its curvature tensor R is selfdual:
R. (R
E
(5.1)
A2 9 A2 ; we dualize R by dualizing either on the first A2 or on the
second.
By the interchange symmetry of R, it doesn't matter which.)
From
(5.1) it follows that the Ricci tensor vanishes, so every selfdual metric is a solution of Einstein's vacuum equations. Recall our comment in §2 that the only real selfdual 2form in Minkowski spacetime is zero.
For the same reason, the only real selfdual metric with
Lorentz signature is flat (i.e.
Minkows~i).
To escape from this trivial
situation, we could either allow g to be complex or require that g be positive definite.
Both these cases have applications: for example, Hspaae (which
is a complex selfdual space  see Newman's lecture) and "g:roavitation.al
instantons" [8] (which are positi.ve definite). Hhat we are after is a theorem relating selfdual metrices to "deformed" twistor spaces with a certain structure.
To see what this structure is, let
us examine the "flat" case of §2 more closely. Let
Xbe
the region of P 3 corresponding to an open ball X in CM.
Pro
position (2.1) implies that
29
A
{a)
X is a holomorphic bundle over P 1 •
Here P1 is the projective version
of the spinspace S~: in other words, P 1 = (S~ {0})/C*. bundle (a) is a domain in bundle L(l) over P1 •
c2 •
Each fibre of the
Xof
Let L denote the pullback to
the line
Let V be the subbundle of the tangent bundle of X
consisting of the vertical vectors, i.e. of the vectors tangent to the fibres of (a).
The second bit of structure we want to single out is
a skew bilinear map
{b)
~:V
x V
+
2
L •
(In the flat case,
~
comes from the
symplectic form on S' .) Theorem 5.1 (i)
There is a 11 correspondence between
sufficiently small selfdual deformations of the metric in X; and
(ii) sufficiently small deformations of
X,
preserving the structure (a) and
(b).
Remarks (1)
Various bits of the proof may be found spread over a number of references,
for example [10,5,6]; the deformation theory of Kodaira is used extensively. An outline of the correspondence will be given below. (2)
The above theorem is a local version: in this context "local" means
virtually the same thing as "sufficiently small".
Conversely, a global
theorem would have to do without the "sufficient smallness". global positive definite theorem.
See [2] for a
The main difference arising in this case
A
is that some extra structure on X has to be specified; in the case of Theorem 5.1, this extra structure is guaranteed by Kodaira's deformation theorems and so it doesn't have to be mentioned explicitly. Let us now describe the correspondence in Theorem 5.1. we are given X with a selfdual metric g.
Suppose first that
Define a Ssurfaae in X to be a
complex 2surface which is totally null and antiselfdual; then let
30
Xbe
the
space of esurfaces in
x.
x is
The selfduality of g guarantees that
a
3dimensional complex manifold (as long as g is sufficiently close to being A
flat) and that the structures (a) and (b) can be defined on X in a natural way.
Xbe
Conversely, let
a deformation satisfying (a) and (b).
We want to
construct X with a selfdual metric: this is done as follows.
X is defined
to be the space of holomorphic sections of the bundle X~ P1 •
A theorem of
Kodaira [6] guarantees (for sufficiently small deformations) that X is a 4dimensional complex manifold. holomorphic section
i
and let V(x) be the bundle of vertical vectors re
stricted to the submanifold
A vector vET
X
a metric on T • X
decompose T
i.
So V(x) is the normal bundle of
in X corresponds to a section~ of V(i).
i
in
X.
We want to define
To do so, it is sufficient (cf. Equations (1.1)  (1.3)) to
into spinspaces
X
S 8 S'
T
X
X
(5.2)
X
and to define symplectic forms £
= V{x) 8 sphere i).
Put W
X
Riemann
Suppose that x E X corresponds to the
X
and £ 1 on S X
X
and S' respectively. X
L(1) (where L(1) isthought of as a bundle on the Thus
W 8 L{l). X
When we look at sections of V(i), it turns out that
r(w) 8 r(L(l)). X
(5.3)
This is proved [5] by establishing that it holds for the flat case and that the various bundles involved are stable under small deformations.
Now define
31
S'
:=
r(W ),
sX
:=
r(L(l)).
X
X
(5.4)
r(v(i)), (5.3) and (5.4) give us the desired decomposition (5.2).
Since T
X
The symplectic form E' on S' is derived from the skew map X
].1:
v
X
v
X
2 L •
+
(5.5)
Restricting (5.5) to x and tensoring with L(1) leads to ].1
X
:
w
X
X
w
X
+
L ( 0)
and taking global sections of this gives us E'
X
S
as required.
X
X
S
X
+
C,
The symplectic form E on S
X
is defined as follows.
Going back
to the flat case, the symplectic form on S* induces a symplectic form E on 0
the space r(L(l)) of crosssections of L(l) over P1
(S*  {O})/C*. 0
This
structure is preserved under deformation (since the base space P1 is preserved) and so gives usE on Sx
= r(L(l)).
This completes the definition of the
metric, which turns out to be selfdual. Finally, a few words about
g~avitationaZ
instantons.
How one defines
these depends on what one wants to use them for, but let us say that they are 4dimensional Riemannian manifolds t.•hich (a)
are complete
(b)
have vanishing Ricci tensor
(c)
are either compact or (in some suitable sense) asymptotically flat.
Some of these spaces will be selfdual and will therefore correspond to deformations of twistor space satisfying some extra conditions.
These extra
conditions are not, as yet, clearly understood and a great deal of work
32
remains to be done on the whole subject of gravitational instantons. What I would like to mention here are some functionals which are closely analogous to the over
Xwhich
Aq
Recall that L denotes the line bundle
of Theorem 4.2.
is the lift of L(l) over P1 •
look for holomorphic sections of Lq.
Let q be a positive integer and
Since f(L(q)) _ Cq+l, we know that (5.6)
(i.e. sections of L(q) over P1 can be lifted to X). A section of Lq over corresponds, in the selfdual spacetime X, to a solution
equation.
This means that
~ €
S*(q) on X, with (ToV)
~
~
=0
of the
X
twisto~
[cf. Equation
(2.2), where the twistor operator ToV is defined for flat spacetime; the definition for curved spacetime is the same]. Now in general the equality sign in (5.6) holds: all sections of Lq come from the base P1 (equivalently, all solutions of the twistor equation are covariantly constant).
Let G be the set of selfdual metrics on X admitting q
a nonconstant solution of the twistor equation (i.e. so that dim r(Lq) > q + 1).
Then we might
conject~e
that all the elements of G
q
arise as the images of functionals A , similar to those of Theorem 4.2, q
mapping solutions of a metrics.
line~
equation into the space of all selfdual
This conjecture is true for q
some of these functionals.
= 1,2:
see [11] for the details of
For q > 2 not much is known.
It is also not yet
known whether there is an analogue of Theorem 4.3, i.e. whether all selfdual gravitational instantons can be generated by such functionals.
33
REFERENCES [1]
M.F. Atiyah, N.J. Hitchin, V.G. Drinfeld and Yu. I. Manin, Phys. Letters 65A (1978), 185187.
[2]
M.F. Atiyah, N.J. Hitchin and I.M. Singer, Selfduality in fourdimensional Riemannian geometry (preprint, Oxford, 1978; to appear in Proc. Roy. Soc.).
[3]
M.F. Atiyah and R.S. Ward, Comm. Math. Phys. 55 (1977), 117124.
[4]
E.F. Corrigan, D.B. Fairlie, R.G. Yates and P. Goddard, Comm. Math. Phys. 58 (1978), 223240.
[5]
W.D. Curtis, D.E. Lerner and F.R. Miller, Some remarks on the nonlinear graviton (to appear in Gen. Rel. Grav.)
[6]
R.O. Hansen, E.T. Newman, R. Penrose and K.P. Tod, The metric and curvature properties of Hspace (preprint, Pittsburgh, 1978; to appear in Proc. Roy. Soc.).
[7]
R. Hartshorne,
Comm. Math. Phys.
[8]
S.W. Hawking,
Phys. Letters
[9]
R. Penrose,
J. Math. Phys.
[10]
R. Penrose,
Gen. Rel. Grav.
[11]
R.S. Ward,
A class of selfdual solutions of Einstein's equations
~
(1978), 115.
~2~
(1977), 8183.
~
(1967), 345366.
l
(1976), 3152.
(preprint, Oxford, 1978; to appear in Proc. Roy. Soc.).
R.S. Ward Merton College Oxford England
R Hartshorne Algebraic vector bundles on projective spaces, with applications to the YangMills equation My purpose in these lectures is to give some idea of recent work in algebraic geometry concerning algebraic vector bundles on projective spaces, and how it is related to the solutions of the YangMills equation.
Since I have recently
written three other articles on this subject, I will confine myself here to a general account of the most striking success in this area, namely how methods of algebraic geometry have led to a "complete" solution in terms of linear algebra of the problem of classification of instantons.
This is only one of
several recent instances of algebraic geometry appearing in connection with some of the differential equations of mathematical physics:
two others are
Hitchin's work on gravitational instantons (reported by Penrose at this conference) in which the rational double points of algebraic surfaces play a central role, and the work of Kri~ever, Mumford, and van Moerbeke relating certain line bundles on an algebraic curve to the solutions of the KortewegDeVries equation [10]. For further details I will refer to the survey article [7] about the connection between vector bundles and instantons, the detailed account [8] of the theory of rank 2 vector bundles on P 3 , and the list of open problems [9], which also contains a long bibliography.
For general background in algebraic
geometry see [6].
Jackiw in his talks at this conference explained why physicists are interested in instantons.
The natural world should best be described in terms of quantum
field theory, but the equations which arise in that theory are too difficult
35
to solve.
The corresponding classical equations give a first approximation.
Then by allowing time to become imaginary, one obtains solutions of the classical equations which are forbidden in real time, and which are thought to reflect certain quantummechanically allowed motions such as "tunnelling" between two zeroenergy states of a physical system.
This leads in particular
to the search for solutions of the classical YangMills equation with imaginary time, i.e., in Euclidean space lR 4 •
For purposes of computation,
the most important solutions will be those with minimal action:
these are
the selfdual (or antiselfdual) solutions, which are called instantons. They are indexed by an integer k
~
0, the "instanton number," and for each k
one expects a continuous family of solutions depending on auxiliary parameters such as "position" and "scale."
The physicists were able to find some
solutions for each k, but could see by a parameter count that they had not found them all. Ward in his talks at this conference (see also [3]) explained how the general ideas of Penrose allow one to transform the study of instantons into a problem of complex analysis. 4
On S
First we compactify
~
4
4 to the 4sphere S •
for each integer k there is just one principal SU(2)bundle with second
Chern class c 2 = k (we will stick to the group SU(2) for simplicity, although completely analogous results hold for all other compact gauge groups.)
A
selfdual solution to the YangMills equation on S4 corresponds to a connection on this SU(2)·bundle whose curvature is selfdual. Now the Penrose transform appears in this case as a mapping w:~+
4
from complex projective 3space to S •
1 3 complex projective line PC: in P 1 •
s4
4 corresponds to a
Each point of S
This complex projective space has a real
structure a on it for which there are no real points, but there are real lines, namely those of the form w1 (x), for xES 4 •
4
Given an SU(2)bundle on S • one can lift it by w to an SU(2)bundle on
~~. which is associated to a certain ~ 2 bundle E.
A connection with self
dual curvature on the original SU(2)bundle gives an almostcomplex structure on E, and the selfdual condition is the integrability condition needed to make E a holomorphic rank 2 vector bundle on ~
3
P~.
s4 •
Since it comes from
it
~2
has a real structure a with a = 1, and clearly the restriction of E to each real line w 1 (x) is trivial. The main result of Atiyah and Ward [3] is that this is an equivalent to give an SU(2)instanton on S4 is equivalent to giving a rank 2
problem:
holomorphic vector bundle Eon P~ whose restriction to each real line n 1 (x) ~
is trivial, and which has ·a real structure a lying over the real structure on
3
P~.
The instanton number k appears as the second Chern class c 2 (E),
while c 1 (E) is necessarily 0. Thus we have arrived at a problem in algebraic geometry.
Part of it is
purely holomorphic, namely the study of holomorphic vector bundles on
3
P~.
An old theorem of Serre states that every holomorphic vector bundle (and more
generally every holomorphic coherent sheaf) on a complex projective space has a unique algebraic structure, so that the theory of holomorphic vector bundles is equivalent to the theory of algebraic vector bundles.
The other
part of the problem, involving the restriction to the real lines and the real structure, can be phrased as a problem in algebraic geometry over approach will be to study the problem over
~
~.
Our
first, and then tack on the real
structure afterwards. The first step is to study the classification of algebraic rank 2 vector 6undles on
P!.
This is a problem which has concerned algebraic geometers
for some time, in particular Horrocks, Barth, Maruyama, and myself.
There
are two numerical invariants, the Chern classes c 1 , c 2 , which can take on all
37
integer values satisfying c 1 c 2
=0
(mod 2).
Having fixed c 1 , c 2 , there are
infinitely many families, depending on larger and larger parameter spaces, of bundles with the given Chern classes.
The parameter spaces are not well
behaved topologically (e.g. they may not be Hausdorff spaces). introduced the concept of a stable vector bundle.
So Mumford
The stable bundles with
given c 1 , c 2 are parameterized by a finite union of quasiprojective varieties, which we call their moduli space.
Also it happens that the bundles
arising from instantons are all stable, so that we have good moduli spaces for them. Fixing c 1
= 0,
c 2 > 0 for simplicity (which includes in particular all the
bundles coming from instantons), we can then speak of the moduli space M(c 2 ) of stable algebraic rank 2 vector bundles on P~ with those Chern classes. Unfortunately the structure of this space is rather complicated and not yet well understood.
It may have several different irreducible components of
different dimensions.
It may not be connected.
irreducible component must have dimension
~
What is known is that every
8c 2  3, and that there is at
least one irreducible component of dimension exactly 8c 2  3 which contains the bundles coming from the (5c 2 + 4)dimensional family of instantons found by the physicists. Here we have been talking about the complex dimension of the complex analytic space M(c 2).
Because of the reality conditions which are imposed
on the bundles corresponding to instantons, the parameter space M'(k) of instantons with instanton
number k will be a union of certain connected
components of the real part of M(k).
Thus it will be a real analytic space
whose real dimension is equal to the complex dimension of M(k).
And indeed
it .is known that the moduli space M'(k) of instantons is a realanalytic manifold of real dimension 8k  3 (this was proved by Atiyah, Hitchin, and
38
S!nger [2] using methods of differential geometry).
However, it is not yet
known (except fork= 1, 2) whether M'(k) is connected.
There are two principal methods for studying rank 2 vector bundles on P 3 • one is via curves, the other via monads. The curve method goes back to Serre (1960), was used by Horrocks (1968), and was apparently rediscovered independently by Barth and Van de Ven (1974) and Grauert and Mulich (1975). idea is this.
It is the main technique used in [8).
The
Tensor the bundle E by a multiple of the Hopf line bundle 0(1)
so that the new bundle E(n) has plenty of global sections. sufficiently general section, then the set of points in P
3
If s is a where s becomes
zero will be a curve Y in P 3 • Conversely given a curve Y together with certain extra data I won't state explicitly here, one can recover the bundle E. Thus the study of vector bundles is reduced to the study of curves in P 3 This is another difficult subject, but a lot is known, and it has a venerable history going back into the 19th century. examples of bundles.
This method leads easily to many
A first series of examples, for every k > 0, can be
constructed by taking the curve Y to be a disjoint union of k + 1 lines in These examples contain already all the bundles corresponding to instantons found by the physicists. k
>
A second series of examples for every
0 can be constructed by taking Y to be an elliptic curve of degree k + 4.
For k ~ 3 these are new, not contained in the first series.
There are other
series after that, corresponding to curves of higher genus, but they cannot be described so explicitly because we do not know the corresponding curves so well.
39
The
ot curves also lends itRelf well to the explicit description
~ethod
of bundles with small values of c 2 • bundles on P
3
with c 1
= 0,
and c 2
Thus for example all stable rank 2
= 1,
2 have been classified, the
corresponding moduli space described, and the moduli space of instantons with k
= 1,
2 also described explicitly.
The other main method of studying bundles on P 3 is by monads.
According
to Webster's International Dictionary, a mopad is a unit or atom in Greek philosophy.
In the philosophy of Leibnitz it is "an individual elementary
being, psychical or spiritual in nature, but constituting the underlying reality of the physical as well."
The word monad was used by Horrocks (1964)
to denote a sequence
F'~ F..JL. F" of vector bundles and morphisms of bundles such that Sa and
B is
monad.
surjective.
= 0,
a is injective,
The bundle kerS/ima is called the homology of the
Horrocks showed, in a letter to Mumford (1971), that every vector
bundle E on P 3 can be obtained as the homology of a monad in which each of the bundles F', F, F" is a direct sum of line bundles. kind of "twosided resolution" of the bundle E.
Thus the monad is a
Although a monad is a more
complicated object than a single bundle E, it is easier to describe explicitly, because the individual bundles F', F, F" are simple, and the maps a,
B can
be represented by matrices.
Thus in principle the use of
monads reduces the study of vector bundles to linear algebra. Barth observed for the case of stable rank 2 bundles E on P 3 with c 1 c2
=k
that if one imposes the extra condition H1 (E(2))
corresponding monad takes on the special form
1.1:0
= 0,
then the
0,
To specify such a monad, one has only to give a and 6, which are (2k + 2) x k and k x (2k + 2) matrices of linear forms in the homogeneous coordinates z 0 , z1 , z 2 , z 3 of the projective space. subject to the conditions that rank a Ba
= 0.
Of course the matrices a, B are
= rank B = k at every point, and that
Barth showed further that this monad was selfdual, so that B is the
transpose of a with respect to an alternating bilinear form J on 0 The extra cohomology condition H1 (E(2))
2k+2
.
0 remained somewhat mysterious
until Atiyah and Hitchin, and independently Drinfeld and Manin, showed that this condition holds for all those bundles coming from instantons, 4
proof is by differential geometry on S
The
To quote their paper [1], it is
because the linear differential equation
(!:,.
1
+ 6R)u = 0 has no global nonzero
solutions, where u is a section of the SU(2}bundle on
s4
corresponding to E,
A is the covariant Laplacian of that bundle, and R > 0 is the scalar curvature of S4 •
Putting these results together shows that every bundle coming from an
instanton can be represented by a monad of the above special form. To clarify this situation further, I would like to quote the form of Horrocks's monad construction used by Drinfeld and Manin (and explained very clearly in their paper [5]), which shows that the cohomological condition 1
H (E(;2)) = 0 mentioned above is essentially equivalent to the possibility of representing E by a special monad.
It is this.
There is an equivalence
of categories between (i)
the category of coherent sheaves F on lP 3 satisfying H0 Q:(m))
0
for m ~ 1
H1 (F(ml}
0
form~
2
H2 Q:(m)) = 0
for m > 2
H3 Q:(m}) = 0
for m > 3
~1
and (ii)
special monads of the form
A® 0(1)
~ B ® 0 J..._ C ® 0(1)
where A, B, C are complex vector spaces. In the case of a bundle E of rank 2, the condition on H0 is impli.ed by 1
1
stability, the condition on H is implied by H (E(2)) = 0, and the conditions 2 3 on H and H follow by Serre duality. The important point about this theorem is that it gives an equivalence of categories.
This means that if the bundle E has any additional structure,
then this additional structure is automatically inherited by the corresponding monad.
In particular, if E has rank 2, then it carries an alternating 2
bilinear form arising from the natural mclp E ® E  A E ~ 0, so the corresponding monad also carries this form.
If E comes from an instanton,
"'
then it has a real structure a, and this also carries over to the monad. This achieves the reduction of the problem of classification of instantons to a problem in linear algebra.
Each SU(2)instanton with instanton number k 3
corresponds to a stable rank 2 vector bundle on Pt which can be represented by a special monad as above. conditions carry over.
The monad is selfdual, and the reality
To describe the monad, we have only to give the map
a, and list the conditions it must satisfy. (2k
+
2)
X
The map a is given by a
k matrix A(z) of linear forms in the homogeneous coordinates zo·
z 1 , z 2 , z 3. of the projective space. (1)
A(z)tJA(z) = 0
for all z
(2)
rank A(z) = k
for all z
(3)
J A(z) = A(az)
It is subject to the conditions
~
0
where J is the (2k + 2) x (2k + 2) skew symmetric a(zO,zl,z2,z3)
matrix(~ ~)and
(zl,z0'z3,z2).
Condition (1) expresses the relation Sa
=0
when S is taken to be the
transpose of a via selfduality of the monad; (2) is the rank condition which makes a injective and
B surjective; (3) is the reality condition.
There is ambiguity in the choice of bases for the bundles on the left and in the middle of the monad.
Thus two matrices A(z) and A'(z) give the same
instanton if and only if A' (z)
S • A(z) • T
where S E Sp(k + 1) and T E GL(k,E).
So the moduli space M'(k) of
instantons can be obtained as the space of all such possible matrices A(z) modulo the action of the group Sp(k + 1) x GL(k,E). Some explicit calculations of the moduli space have been made from this point of view by Rawnsley, Jones, and by Christ in his talk at this conference (see [4]).
While this matrix formulation has been hailed by some
as a "complete solution" of the problem of classifying instantons, there are still many questions left unanswered.
The calculations quickly become very
complicated, so that for example one still does not have an answer to the (deceptively) simple question whether the moduli space M'(k) of instantons is connected for k
>
3.
References 1
K. F.
Atiyah, N. J. Hitchin, V. G. Drinfeld, Ju. I. Manin, Construction of instantons, Physics Letters 65A (1978) 185187. 
2
M. F. Atiyah, N. J. Hitchin, I. M. Singer, Deformations of instantons, Proc. Nat. Acad. Sci. USA, 74 (1977) 26622663.
3
M. F. Atiyah, R, S, Ward,
Instantons and algebraic geometry, Comm. Math. Phys. 55 (1977) 117124.
4
N. H. Christ, E. J. Weinberg, N. K. Stanton, General selfdual YangMills solutions (preprint)
5
V. G. Drinfeld, Ju. I. Manin,
Instantons and sheaves on ,F 3 (preprint)
6
R. Hartshorne,
Algebraic Geometry, Graduate Texts in Math 52, SpringerVerlag, New York (1977rxvi + 496 pp.
7
R. Hartshorne,
Stable vector bundles and instantons, comm: Math. Phys. 59 (1978) 115.
8
R. Hartshorne,
Stable vector bundles of rank 2 on P 3 , Math. Ann. (to appear)
9
R. Hartshorne,
10
D. Mumford,
Algebraic vector bundles on projective spaces: a problem list. Topology (to appear). An algebrageometrical construction of commuting operators and of solutions to the Toda lattice equation, KortewegD~ Vries equation and related nonlinear equations, Kyoto Conference (to appear)
ROBIN HARSHORNE Department of Mathematics University of California Berkeley, CA 94720
N H Christ Selfdual YangMills solutions
Let us consider the application of the HorrocksBarth construction to the problem of finding selfdual Euclidean YangMills solutions, recently developed by Atiyah,
Hitch~n,
Drinfeld and Manin [1].
In this talk I should
like to show explicitly how the construction of these authors can be reduced to the problem of solving a nonlinear equation for a matrix of quaternions. As we will see, this quaternionic formulation has a very simple and direct connection with the original, physical gauge theory and yields, for example, all SU(2) solutions with Pontryagin index k=3 and Sp(n) solutions with k
One begins with the monad
w*
(1)
k 2k+2n are complex vector spaces of dimension k and 2k+2n where W and V a respectively while the sum M z
a
. k 2k+2n is an inJection of W into V which
depends linearly on the four homogeneous coordinates Furthermore
v2k+2n
{z~
l
~P 3 •
is identified with its dual by introducing a symplectic
bilinear form 0k+n defined on V2k+2nthus allowing the dual mapping, (Ma z a ) * , to act on v 2k+ 2n. point z in
~p 3
The fiber of the resulting rank two vector bundle over the
is given by (2)
In order to obtain a bundle corresponding to a symplectic gauge group we ~ 2k+2n must introduce conjugate linear involutions ow and oV on wr and V obeying
(3)
and demand
(4) where a is the mapping on ¢F 3 (5)
We can choose a basis for
if
and
v2k+ 2n
with respect to which ow is simply
complex conjugation while oV is complex conjugation followed by transformation with the 2(k+n) x 2(k+n) matrix
0 1 0 0
1 0 0 0
0 0 0 1
0 0 1 0
'
(6)
'' 0
1 1 0
Thus
our bundle has a conjugation oE which lies over the operator a of
Eq. (5) on aF 3 and a symplectic, bilinear form induced by the form nk+n. The connection between the conjugation oE and our search for symplectic gauge fields can be seen by referring to Ward's construction [2].
Recall
that a symplectic matrix S in Sp(n) is a 2n x 2n unitary matrix (7)
which preserves the symplectic form On: (8) ~6
combining these equations one obtains, (9)
where x
£
S is the complex conjugate of the matrix S.
Thus for real points
~ 4 a symplectic connection matrix A~(x) will obey Eq.
(9) and the
transition functions obtained in the Ward construction,
Gij(z) = P { exp
Jxji
(z)
x (z)
(10)
A (x)dx } 1.1
1.1
will be compatible with a conjugation aE similar to that induced on E2 by Eq. (4): Gij (a(z))Q
n
The right hand side of Eq. (10) is the usual pathordered integral.
(11)
The
points xi(z) and xj(z) lie on the antiself dual plane determined by the point z in G:lP 3 ,
(12)
(where {x }0 1.1
<:3 are the four components of the point x), and are chosen
~~
to obey x(z)
x(a(z)).
(13)
Finally, the condition (8) implies in a more straight forward way the existence of a symplectic bilinear form on the resulting bundle:
(14) 4,7
The preceeding arguments can be easily transcribed to the orthogonal groups O(n), in which case Eq. (8) and the reality condition (11) apply without the matrix Q, construction
Consequently in the Atiyah, Hitchen, Drinfeld, Manin
2
av = +1
so that one must choose
consistent with the property a
2
=
aw2 = 1
if Eq. (4) is to be
1, implied by Eq. (5).
The unitary groups,
U(n), are somewhat different since the fundamental representation of U(n), n>4 is inequivalent to the complex conjugate representation.
Instead, for
U(n), n>2, one can use Eq. (7) and introduce a conjugate linear mapping which relates the bundle E2 with its dual E* 2• Let us now develop a simple representation for the mapping Maz a of Eq. (1) and the implied requirement (15)
If we introduce a spinor notation (16)
so that Eq. (12) becomes (17)
then Maz
a
can be written as the matrix (18)
Using the basis previously adopted, the reality condition (4) becomes
nB
" A
= BA,
nc
" A
where for a spinor
48
~A
 c~A
(19)
C)(_::) A Thus if B2i+al,m and
(20)
A
c2i+al,m
for fixed 1 and m are viewed as 2x2 matrices
with indicies a and A, they belong to a 2x2 representation of real quaternions
1
=(~ ~)
Consequently we will view B,C, and M(x)
(21)
~·
k{:
(~ ~
j
G01 ~
i
= BCx
as (n+k)xk matrices of
quaternions where xis the quaternion represented by the matrix in Eq. (12). The requirement (15) can be written (22a)
0
or (22b) t
Where M (x) is the transpose of the matrix whose elements are the quaternionic conjugates of those in M(x): (M(x)t)im
(23) A
i, (j)
t
A
= j
A
and (k)
t
A
= k.
The indicies A and B
in Eq. (22b) again refer to the 2x2 matrix description, Eq. (21), of the quaternions. R(x)
The validity of Eq. (22b) for all n requires that the matrix Mt(x)M(x)
(24)
be real.
Thus the problem of finding selfdual
Sp(n) gauge fields has been
reduced to that of finding a (k+n)xk matrix of quaternions M(x) of the form B Cx such that the product Mt(x)M(x) is a real, nonsingular matrix. This matrix formulation has the added practical advantage that the gauge field, of physical interest, can be directly constructed in terms of it
[3,4].
First define a (k+n)xn quaternionic matrix N(x) obeying t
N (x)N(x) t
N (x)M(x)
I
(25a)
o.
(25b)
Then the gauge field corresponding to the original bundle E2 is given by (26)
If the quaternionic elements of A (x) are replaced by 2x2 matrices as in ~
Eq. (21) and the result viewed as a 2nx2n matrix, one obtains a skewhermitian gauge field obeying Eq. (8), i.e., a symplectic gauge field.
It
is easy to show directly that the field strength corresponding to A (x} is ~
selfdual.
Finally, and perhaps most remarkably, the matrix N(x) can be
used to find a solution [3,4] to the equation
DD ~
~
= ~ 4 (xy),
~(x,y)
(27)
where D
~
a + A (x) • = aX~ ~
(28)
The solution to Eq (27) is simply
~(x,y)
1 = 4n 2
50
Nt(x)N(x) (xy) 2
(29)
As a last topic we will discuss two cases in which this simple quaternonic formulation allows the explicit construction of new, selfdual solutions. We can choose a canonical form for the matrix C in which the first n rows vanish and the last k rows make up a kXk unit matrix.
With this choice of C
the condition that Mt(x)M(x) be real, requires that BtB be real and that the last k rows of B make up a symmetric matrix.
If we let {qij}l~i~n,l~~ be
the matrix formed by the first n rows of B and {b i.} 1 <1., j
the remaining k rows, these conditions are that bij


= bji
and that
(30)
be real.
Thus the qij represent 4nk real parameters while the symmetric
represent 2k(k+l).
3
The reality of the kxk matrix (30) imposes lk(k1)
conditions leaving 4nk + ~(k+7) parameters.
However, the selfdual
solution (26) is not altered and our canonical form for C and the condition (30) preserved if we multiply M(x) on the right by a kxk orthogonal matrix 0 (k(k1)/2 parameters) and on the left by a (n+k) x (nxk) matrix S made up 1
of an nxn symplectic matrix G and 0
:
.G :1) (n(2n+l) parameters).
(31) Thus there remain 4(n+l)kn(2n+l) meaningful
parametersthe numbe= given previously [5] for general Sp(n) selfdual solutions when k>n.
When
k~n,
the above counting must be modified since the
nxn symplectic matrix G acts on only kxn independent columns of qij' Consequently q
will be unaffected by a Sp(nk) subgroup of Sp(n), yielding ij an additional (nk)(2n2k+l) parameters for a total of 2k 2 + 3k.
51
In fact, in the case
n~k,
a general solution to the nonlinear requirement
that the expression (30) be real can be written down [6].
The elements
bij = bji are first chosen arbitrarily (2k(k+l) parameters).
Next the
freedom implied by the Sp(n) transformations G in Eq. (31) is used to choose qil
=0
for i > 1 and q11 real and positive.
Next, this symplectic freedom
is exploited tochooseqi 2 = 0 for i>2 and q 22 real and positive. of Eq. (30) for i=l, j=2 then fixes the
qua~ernionic
The reality
part of q 12 •
This
process can be continued inductively producing a matrix qij which vanishes for i>j, is real and positive for i=j, and has a determined quaternionic, part for i
Thus the reality of the expression (30) is quaranteed and qij
contains k(k+l)/2 real parameters for a total of
25
2 3 k + 2 kconsistent with
out previous result since the O(k) redundancy has not been removed. similar result can also be obtained for O(n),
4n~k,
and SU(n),
A
2n~k.
For fixed n and general k it is not yet known how to write down the general solution.
However, the 5k parameter solutions of 't Hooft can be
easily obtained in this language. (the square root of the i
th
Choose the qli to be real and positive
instanton scale) and bij to be diagonal (bii is
the position of the i th instanton). A second case [4] in which an explicit general form for Bij can be found consistent with the reality of the quantity (30) is the case of Sp(l)=SU(2) and k=3.
We first use the Sp(l) x 0(3) symmetry of our formulation to choose
q11 real and to diagonalize the real part of bij'
The remaining components
of bij are then treated as the 21 = Sk3 physical parameters and the qli' l~i~~
52
fixed by the requirement that the quantity (30) be real.
The result is
re[w3w2 (w1w3w3w1 )]
(32)
2lw2w3w2w212 re[w 3w2 (w1w2w 2w1 )] 2lw2w3w3w212 where the quaternions wk are defined by
(33)
In conclusion, it appears that this very explicit quaternionic description of the Horrocks, BarthAtiyah, Hitchin, Drinfeld, Manin construction is a very powerful one, both allowing the construction of new solutions and making very
accessibl~ ~uantities
and the isospin
~Green's
of physical interest such as the gauge field function
A~(x)
~(x,y).
References 1
V. G. Drinfeld and Yu. I. Manin, Funkcional Anal. i. Prilozen (to appear); V. G. Drinfeld and Yu. I. Manin, Uspehi Mat. Nauk· (to appear); M. F. Atiyah, N. J. Hitchin, V. G. Drinfeld and Yu. I. Manin, Phys. Lett., 65A, 285(1978).
2
R. S. Ward,
3
E. J. Corrigan, D. B. Fairlie, S. Templeton and P. Goddard, A. Green's Function for the General SelfDual Gauge Field, preprint.
4
N. H. Christ, E. J. Weinberg and N. K. Stanton, Columbia University preprint, # CUTP119.
5
C. W. Bernard, N. H. Christ, A. H. Guth and E. J. Weinberg, Phys. Rev. Dl6, 2967(1977) and reference therein.
Phys. Lett., 61A, 81(1977).
53
6
V. G. Drinfeld, and Yu. I. Manin, A Description of Instantons, preprint.
Acknowledgment This research was supported in part by the United States Department of Energy.
N H CHRIST Department of Physics Columbia University New York, New York 10027
R Penrose on the twistor descriptions of massless fields 1.
TWISTOR FUNCTIONS AND SHEAF COHOMOLOGY
* a
Recall [1,2] the following properties of a twistor function f(Z ), to be used for generating a zero restmass field ~A' ... L'' or ~A ... K' by means of a contour integral, ~ I I = A ••• L
~A
(i)
•••
K =
1
(211i)
2
f
1ja 2 A
(211i)
aw
11
11
A' ••• L'
. . . Ka
aw
f(Za)d211,
a2
f (Z ) d 11:
The function f is holomorphic on some domain V of twistor space not invarianl under SU(2,2) (nor under the Poincare group, nor the Lorentz group).
(ii)
There is a "gauge" freedom G whereby
where h± is holomorphic on some extended domain V±(~V) of twistor space in which the contour y can be deformed to a point (to the "left" in
*Taken
v
and to the "right" in v+)
(with minor corrections) from Twistor Newsletter 2 (10 June, 1976).
55
contour y
Domain
(iii)
V UV+
(e.g. R +) symbolically represented as S2 •
G depends on the location of y (and y is not invariant under SU(2,2)
nor under the Poincare group nor the Lorentz groupso G is not invariant either}. (lv}
56
By invoking G, then
~oving
y, then invoking a new G,
~oving
y again,
,
......,
'
'

~
...
we can obtain a whole family of equivalent twistor functions, all giving the same field, the entire family being invariant under SU(2,2). (v)
This seems a little nebulous, and, for example, how do we add two such families (to give ~A' ••• L' + ~A' ••• L'), etc. etc?
A new viewpoint concerning twistor functions has been gradually emerging which makes good mathematical sense of all this:
a twistor function is
really to be viewed as a representative function (or cocycle) defining an element of a sheaf cohomology group.
Now, the twistor theorist, when
attacked by a purist for shoddiness in the domains, can counterattack armed with his sheaf! Thumbnail sketch of (relevant) sheaf cohomology theory First let us recall how ordinary (Cech) cohomology works. (a Hausdorff paracompact topological space, say). f.
~nite system of open sets Ui.
Let X be a space
Cover X with a locally
We define a cochain (with respect
to this
57
covering) with coefficients in an additive abelian group 1l , the reals :R,
~.
assigned to the various Ui and their non
fi assigned to Ui; fij assigned to Ui n Uj; fijk
empty intersections:
assigned to ui n uj n uk ••• , and fij
=
f[i ••• R.]"
= fji'
=
(fl,f2,f3, ••• )
1 cochain 13
=
(fl2'f23'fl3'""")
2 cochain y
=
(fl23'fl24'""")
u1
(where
n
= ••• ,
fijk
... '
Then
(l
0 cochain
(say, the integers
or the complex field CC) in terms of a collection of
elements fi, fij' fijk' ••• E
i.e. fi ••• R.
~
u2 , u2 n u3 , u1
n
u3 ,
••• ,U1 n U2 n
u3 , u1
n
u2 n
u 4 , ••• are
the nonempty intersections of Ui's). Define coboundary operator c5 as follows:
"f
123
II
"f
124
II
etc. (where, again, U1 n U2 , ••• are the nonempty intersections). o2
0.
some 13.
We call Y a cocycle if cSy Define the
H~U
}
(X,~)
i
Note:
p H{U } (X, I&), i
th E== cohomology
= 0;
Then we have
we call Y a coboundary if Y
=
~by:
(additive group)/(additive group ) of pcocycles of pcoboundaries • as defined, depends on the covering {Ui}.
What we should
do, to define HP(X,~), is to take the appropriate "limit" of all these
58
c513 for
H{~.}(X,G)
for finer and finer coverings {Ui} of X.
However (for X suitably
1
nonpathological) we can always settle on a particular "sufficiently fine" covering {Ui} where, in effect, there is no "relevant topology" left on each Ui or intersections thereof (i.e. all the Hp (Ui n ••• n Uk,G) vanish for all p > 0although, as it stands, this is somewhat unhelpful because a direct limit is then already involved in the definition of "sufficiently fine"). Then this H{~ }(X,G) =
Hp(X,G).
I shall henceforth assume that such a
i
"sufficiently" fine covering has been taken, and that it is countable and locally finite. Now what does this definition have to do with the familiar "dual" relation to ordinary homology H (X,G)? p
How does Y assign values (elements of G) in a
linear fashion to pcycles in X, where Y is some element of HP(X,G)? Intuitively: Add together the fij for shaded regions (entered by 1cycle); similarly for higher dimensional pcycles. (A complete rigorous discussion can also be given.) ~t~~~1cycle curve)
If Y is defined by (f 12 ,f 23 , ••• ), then y(K) (correctly signed) f's on regions entered by
= f 42 +
K
(closed
f 23 + f 31 + ...
K.
59
What about sheaf cohomology though? ization of the above. slightly.
It's really a rather natural general
But first we must rephrase the concept of a cochain
Rather than thinking of fi as simply an element of <& "assigned" to
Ui, and fij "assigned" to Uij' etc., we think of fi as a function defined on Ui which happens, in the above, to take this constant value "fi" E <&, and we think of fij as a constant function on Ui n Uj with values in <&, etc. (This has the incidental advantage that the requirement Ui n Uj unnecessary.)
" This is still ordinary (Cech) cohomology.
tion to sheaf cohomology is now easily made: are now not required to be constant.
0 is now
~
But the generaliza
the functions fi, fij' fijk'"""
(In fact, we could even allow that the
additive group <&, in which the values of the f's reside, may vary from point to point in X, or, indeed, the f's need not really be "functions" in the ordinary sense at allbut such situations will not be considered here.)
We
may require that the f's be restricted in some wayin particular, for the purposes of the present applications we shall often require the f's to be holomorphic (with, here, <& =
~)
and X a complex manifoldor, we may consider
other related classes of functions. So, what's a sheaf?
Actually, I shan't even bother with a formal defini
tion (which can be found in, for example, the references [3,4,5]).
The
v
essential point is that a sheaf is so defined that the Cech cohomology works just as well as before. each open set U c X.
In fact, a sheaf S defines an additive group <&U for
For example, <&U might be the additive group of all
holomorphic functions on U (taking X to be a complex manifold).
In this case
we get the sheaf, denoted 0, of germs of holomorphic functions on X. Slightly more generally, we might consider "twisted" holomorphic functions, i.e. functions whose values are not just ordinary complex numbers, but taken in some complex line bundle over X (think of "spinweighted" functions, for
60
eKample).
An important example of such a twisted function would arise if X
were taken to be projective twistor space
P~
(or a suitable portion
thereof) and the functions considered were to be homogeneous (and holomorphic) of some fixed degree n in the twistor variable. take
Gu
For each open set U C X we
to consist of all such twisted functions on U, and the resulting
sheaf denoted O(n), is called the "sheaf of germs of holomorphic functions twisted by n" (on X).
More generally we might consider functions whose
values lie in some vector bundle B over X (e.g. we might consider tensor fields on X) and GU would consist of the crosssections of the portion of B lying above U:
I
elements of
'u
tt..::.7
I
~
lf~1
I
B
,..1k.! ~ I I
I
1
I
X
Cochains are defined as before (with fiE <&U , fij E <&U nU , ••• ) and the i
coboundary operator
o,
just as before.
i
j
Then we obtain the pth cohomology
&roup of X, with coefficients in the sheaf S, as (pcochains with coefficients in
)/(pcoboundaries with) coefficients in s .
s;
61
As before, we would need to take the appropriate limit for finer and finer coverings {Ui} of X, but we can settle on one "sufficiently fine" covering if desired.
Provided S is what is called a coherent analytic sheaf (and we
are interested primarily in such sheaveslocally defined by n holomorphic functions factored out, if desired, by a set of s holomorphic relations), then "sufficiently fine" can be taken to mean that each of ui, ui n uj, Ui n Uj n Uk'''' is a Stein manifold [4] (and it is sufficient just to specify that each Ui is Stein).
In effect, a Stein manifold is a holomor
phically convex open subset of ~n (or a domain of holomorphy).
= 0,
Stein and S coherent, then HP(x,S)
if p
>
0.
Note:
If X is
0 and O(n) are
coherent. 1
Twistor functions as elements of H (X,O(n)) Let X be some suitable portion of projective twistor space
~IT,
say some
neighborhood of a line in FIT (corresponding to some neighborhood of a point
+
===r
in Minkowski space) , or say F 'Ir , or P 'Ir •
Suppose we can cover X with two
sets U1 ,u 2 (each open in X) such that every projective line L in X meets ul n u2 in an annular region and where ul n u2 corresponds to the domain of definition of some twistor function f(Za), homogeneous of degree n in the twister Zo.. X
P'Ir
62
Then f
=
so flZ by itself defines a 1cochain clearly
n
f 12 is a twisted function on U1
oa
0, so
=
a is
a cocycle.
a,
U2 •
There are no other Ui
n
Uj's,
with coefficients in O(n), for X.
The 1coboundaries, for this covering,
are functions of the form h2hl' where h2 is holomorphic on u2 and hl' on ul.
that the "equivalence" between twister functions under the "gauge" freedom G that we started out with is just the normal cohomological equivalence between 1cochains
a, a•
with
This suggests that we view the twistor function f as
a=
(h1 ,h 2 ).
that their difference be a coboundary:
a•  a = oa'
really defining us an element of H1 (X,O(n)). But this is all with respect to a {U 1 ,U 2 }.
particul~r
Is this covering "fine" enough?
problem here.
covering of X, namely by
There actually is a technical
We cannot (normally) arrange that u 1 and U2 are Stein mani
folds (and in those exceptional cases when we can so arrange this, we would lose the invariance properties that we are striving for). fact, that this problem is not serious.
It turns out, in
One can show by direct construction
(using the inverse twister functionin the cases n < 2 at least, but probably in all cases) that for any given (analytic, positive frequency) field such a covering by two such sets ul,u2 is sufficient in the case X= lPTt.
Note that though X is invariant under SU(2,2), in this case, the
covering is not.
1 However, the cohomology group H (X,O(n)) is invariant.
1 Let us illustrate this by adding two elements of H (X,O(n)) one of which is defined by a twister function f, with respect to the covering {U 1 ,U 2 }, and the other by f, with respect to {U 1 ,U 2 }, the second covering being a rotated Version of the first.
Schematically:
+
We define a representative cochain for the sum by taking the common refine~
::
ment of both coverings.
(Denote U1
=
~
A
= U1 n
U1 n U1 , U2 A
~
~
=
U2 n U1 ,
~
U4 = U2 n U2 to give the refined covering {U1 ,u 2 ,u3 ,U 4 }.)
..
~
U2 , U3
The 1cocycle
a+ a = a is (f,f,ff,f+f,f,f). Because of the "direct construction" argument mentioned above, this 1cocycle will be cohomologous to (i.e. differing by a coboundary from) a cocycle of the form (O,f,f,f,f,O), so we can refer it back to the original covering if desired (although with ul and u2 perhaps reduced slightly in size). But we need not do so if we prefer not to.
We have a generalization of
the concept of a twistor function, namely as a collection of functions on portions of twistor space, defining a 1cocycle with respect to some covering.
We can actually use such a cocycle directly, obtaining the re
quired spacetime field by means of a. branched contour integral. just illustrate this with an example.
I shall
Suppose that X is covered by 3 open
sets U1 ,U2 ,u3 , where on a pLojective line L in X we get the picture: 6~
u
u
u
u
L
The 1cocycle
a is
(f 12 ,f 13 ,f 23 ).
To get the
field~
•••. we perform the sum
of three contour integrals with connnon endpoints (in ul n u2 n U3): (2Tri)
2
~
•.. • =
•••• fl2d
J 'Y12
2
z+
J 'Yl3
•••• f13d
2
z+
J
•••• f23d
2
z
'Y23
It is a simple matter to check that the cocycle condition f 12  f 13 + f 23 = 0 (in U1 n u 2 n U3 ) ensures that the contour's endpoints can be moved Without affecting the result.
This easily generalizes to coverings with N open sets.
Some applications (where knowledge of sheaf cohomology theory would be helpful). (a)
Charge integrality in the "twisted photon" description as opposed to
vanishing charge in the description
(f homogeneous of degree 0). We have [3] the exact sequence 0 >7l x21lii> O expi> O* i> O,
where 0* refers to nonzero holomorphic functions taken multiplicatively, from which we derive [3,4,5] the long exact sequence: 1
••• i> H (X, 7l) i> H1 (X,0) > H1 (X,0*) > H2 (X,7l) > ...
+
ordinary integer 1st cohomology Choose X as a region in
~
••
+j  a
=
P~
a
ClW ClW
f
+
"twisted photon 11
+
ordinary integer 2nd cohomology
corresponding to a small spacetime tube
surrounding a charge worldline. charge line
spacetime
corresponding to X
66
Then the topology of X is
s2
X
s2
X
R2 • so H1 (X,7l)
1
= 0,
and H2 (X,7l) ,;;;, 7l Ea 7l.
1
Thus the space H (X,O*) effectively contains H (X,O), and is strictly larger than H1 (X,0) if the map to 7l Ea 7l is not simply to the zero element.
The
image in the first 7l is always zero, but the image in the second 7l is the value of the charge. reference [6].)
(This much follows, for example, from examination of 1
From this we see that the H (X,O*) description only works
if the charge has integer value (i.e. lies in 7l) whereas the H1 (X,O) description only works if the charge value is zero.
(In contrast to this,
it may be remarked that there is no restriction on the charge value if the Cl
4homogeneity f(Z ) description
q, A'B' is used, this curresponding to H1 (X,0(4)).) (b)
How to grow (or at least halfgrow) new twisters from old.
There is another way of representing elements of HP(X,O), namely by taking the additive group of all 3closed (O,p)forms modulo the additive group of all aexact (O,p)forms [3,5].
By a (q,p)form is meant a (q+p)form on X
(not generally holomorphic) having q holomorphic differentials and p antiholomorphic differentials:

a(l
k
Then oel = ··· dz a'Zk
A dz
i1
_jP
A••• Adz
vanishing is the condition for
Cl

is a aexact (q,p+l)form whose
to be aclosed.
{We have a 2
that the aclosed (0,0)forms are holomorphic functions.
= 0.)
Note
We can also "twist"
such forms or functions, as before.
67
Applying this to H1 (X,O(n)) we obtain the result that the helicity t<n2) massless fields can be described in terms of forms ga(z• ,Z.)dZa on X, modulo forms
ah(z·_• z.) aza
geneous of degree n in
dZa, where
za
~

az 6
=
~
the functions ga being homo
_ ,
aza
and of degree 1 in
Za
(to balance the dZa).
In
order that the form gadZa be actually defined on the projective space ( appropriately twisted) we require also gaZa
= 0.
The massless field
integral now becomes
q,
A ••• D
1 =::2
(2Ui}
fa A
aw
...
a
y
.
D g (Z ,Z.)dZY
aw
or
(j>A' ••• D'
(See also N. M. J. Woodhouse's article in Twistor Newsletter 2 for further details [7].) There is presumably much freedom in the choice of gY.
I shall suppose
that its singularities can be arranged so that the contour in the above integrals (which is initially
s2 ,
i.e., over the entire projective line
corresponding to the spacetime point in question) may be deformed, freeing za from line I.
za
in the process, until
Writing Wa
=
wa (=Za
= "freed"
Za )
passes through the
Ia 6x6 , we get
The function ga can be taken to be holomorphic in Za and W , with W ga a
Restricting to Wa
68
a
0 whence
0
gaiaB = qXaiaB for some q(Za,IaBXB) homogeneous of degrees (n,2).
Thus,
the massless field integral can be written
which is the standard form for a 2twistor integral. a new twistor out of thin air!
Thus we have conjured
This new twistor Xa is still only half
grown, though, since it appears as IaBX B•
This can be stated as laB~ = 0 axB 2
which implies the massless relation I
= 0. zPxaiaB a q pa azaaxB
aa aB aclosed property ~  ~ = 0 now disappears on a~
a~
In fact, the
w a
it is incorporated in laB~= 0 and Xa ~ = 2q (homogeneity relation). axB axa Acknowledgment I am grateful to George Sparling, Andrew Hodges, and Nick Woodhouse for many helpful discussions and, most particularly, to Michael Atiyah for explaining sheaf cohomology theory so clearly to us and for many insightful remarks. 2.
THE BACKHANDED PHOTON (or, to a cricketer, THE GOOGLY PHOTON)
*
Current twistor dogma presents the following picture of things: (i)
The wave function of a particle or system of particles is to be given by a "twistor function" lji(Xa, ••• ,Za), which is holomorphic in all arguments Xa, ••• ,Za possibly with the modification that some (or all) of these should be dual twistors W , ••• instead (or, conceivably, a
Hughstontype multitwistors xaBY, ••• , etc.). *Taken from Twistor Newsletter 3 (December, 1976).
(ii) More correctly, ~,
say~
~
is (presumably) an element of a sheaf cohomology
k
E H (Q,O), for some suitable open set (or closure of an
open set) Q C 'll' x 'll' x ••• x 'll' say ('ll' = twistor space), 0 being the sheaf of (germs of) holomorphic functions on Q (see previous section). (iii)When a particle interacts it does so by deforming the structure of twistor space, or of Q C 'll' x ••• x 'll',
whereby
~
now plays an active
role in determining the precise deformation in question, thence affecting the behavior of other twistor functions defined on Q. The above is vague in various respects. strongly suggested by two examples: twisted photon [9].
But something like (iii) is
the nonlinear graviton [8] and the
(In fact, for gravitons, (iii) is strictly true only
in the weakfield limitunless some kind of nonlinear sheaf theory can be evoked.)
In the standard procedures for performing deformations (infini
tesimal ones at least) it is only H11 s that are involved.
Sparling has
suggested that new procedures for performing deformations may be needed in order to be able to use Hk's (k.> 1) in an active way.
For the moment, I
prefer the more modest idea that single particles are to be described by
Hl, s, these being the things that interact directly with other particles. H2, s, H3' s, etc. would play roles in describing manyparticle states, with k
elements of suitable H (Q,O) groups defining kparticle states. supply a possible answer to a question posed to me by Feynman:
This could how do you
know whether ~(Xa,Ya,Za) describes a hadron, or three massless particles, or a lepton and a massless particle? 2
Elements of H1 (Q,0), H3 (Q,0), and
H (Q,O) are really quite different kinds of animals, after all!
But the
detailed implementation of this idea has proved elusivenot the least problem being to understand "twistorially" why the spinstatistics relation should hold for manyparticle states.
70
Perhaps the most primitive problem, in trying to push forward with (iii), lies in the fact that a twisted photon is only half a photon, namely the lefthanded half.
Thus, the Ward(Sparling) construction gives a deformation
of Q = ~+ starting from a twister function f(Za), homogeneous of degree zero (f E H 1 (~+,0) or f E H 1 (J>~+,O(O))),
to give a lefthanded photon.
Of course a righthanded photon can be produced by using f(Wa), of degree zero.
But such would be to defeat the purpose of the economy of the twister
description (i).
Simply changing the helicity quantum number (or any other
quantum number) should not involve us in changing the space Q over which the twister function is defined.
Thus maintaining the general programme
(i)  (iii) seems to lead us to the view that some form of deformation of
+ must be possible, which effectively encodes the information provided by an element f E H1 (~+ ,0), when f is homogeneous of degree 4 (i.e., in ~
effect, f E H1 (P~+,0(4)).
My suggestion for this, which relates closely
to an earlier proposal due to Sparling and Ward [10], is as follows. Consider, first, the standard Ward twisted photon.
This is obtained by
deforming the bundle (see figure) where the base space remains the undeformed
~+
71
P~+ and the fibre¢ {0}.
Furthermore, the Euler operator
Za __a_ =: TZ remains globally defined.
aza
the forms IZ:= VZ:=
t
f4 EaBy 6 dZa ~
However, with nontrivial twisting,
dZB ~ dZY ~ dz 6 and
EaBy 6 ZadzB ~ dZY ~ dZ 6 are not well defined; indeed, globally
defined analogues of these forms do not exist at all on a nontrivially twisted Ward photon. flat ~,
There is a good reason for this.
In the case of
there is a canonical way of representing a twistor Za (up to the
fourfold ambiguity ±Za, ±iZa) in terms of 1P11', namely as the projective twistor 1l in lP ~, lifts into
~
together with a 3form at 1l in P
~.
This 3form
along the fibre over 1l; where it agrees with VZ defines us
the point Za (or iZa, or za, or iZa).
Thus, if Vz is canonically known in
the bundle, then the complete bundle structure is determined uniquely in terms of P
~;
therefore no twisting in the "photon" can occur.
(This
construction is the analogue of how one defines a spinor KA (up to sign) in terms of the celestial sphere.)
Furthermore, if a Vz exists globally in a bundle over P be unique up to an overall constant factor.
72
~+ ,
then it must
Suppose, generally, we have a bundle (or holomorphic fibration) just locally, which is just a complex 4space over a complex 3space with fibre a complex 1space.
Then all we know locally in the 4space is a direction
,
rI
'
...,

I
 
C P'll'
.......
3space
Vz
field 6Z (!foliation).
To know
in the 4space would be to know rather
more structure than 6Z.
Being a 3form,
Vz
is orthogonal to (i.e.,
annihilates) precisely one complex direction and so serves to define 6Z. (I shall always assume
Vz
to be restricted to be orthogonal to 6Z, so as
Vz
to qualify to be a "VZ.")
Now
dVZ = 4DZ.
and DZ together define the Euler operator TZ,
Furthermore
roughly speaking by "Tz
Vz
= Vz
i
also defines a volume 4form DZ by
IJIZ," or more precisely by
Conversely, this relation shows us that (Tz,IDZ).
VZ
Vz = TZ
_J
IDZ.
is determined by the pair
To know one or the other of Tz• IDZ is, by itself, not sufficient
to determine
Vz,
but the two together are equivalent to
Vz.
Furthermore,
assuming that TZ is to point along 6Z (where 6Z is given), there is precisely as much information in TZ locally as there is in mz.
Each provides us with
a kind of local scaling, but it is a "homogeneity degree 4" scaling for mz
73
and a "homogeneity degree 0" scaling for TZ.
In a sense D>Z and TZ seem to
be sorts of duals to one another. We can regard the Ward photon [9] as arising when we retain only the TZ scaling and throw out D>Z. while throwing out TZ.
Let us try to do the "dual" thing and retain D>Z
I shall proceed in a fairly explicit way, assuming
that two "coordinate" patches are given, where a standard twistor description ~
~
is given in each patch, with X on the left·and Z on the right.
I am
I
I I
I
~X~ I
I
DX=IDZ
• z~ I
I
I I I
!
(
.
.
)
X
)
~
assuming there is no monkeybusiness in the base space, so X assumed on the overlap region.
~
Z may be
(The fibres in each half are given when X~
or Z~ is held constant up to proportionality.) D>X
~
The hypothesis is now
= D>Z on the overlap (instead of Ward's TX = TZ)' i.e., dVX = dVZ,
i.e., d{VXVZ}
d{f(Z~)VZ}
= 0,
= 0.
But
ox = oz,
so Vx
·x
Vz.
Put VX
~
= {1 + f(Z ) }VZ; then
which holds iff f(Z~) is homogeneous of degree 4.
The
transition relation is then (A)
74
This appears to be a particular case of a SparlingWard [10] construction given earlier for massless fields of any helicity. We run into trouble owing to branch points arising from the fourth root unless we exclude an extended region about the origin of each fibre. things are okay near like~
~
on the fibres.
But
Thus, I envisage the fibres as being
with some bounded (probably connected) region removed (i.e.,
remove
1
...__________,1 p 'Jl'
biholomorphic to ~  {z : I~ < 1}). patching, but related by TX
Note that TZ is not preserved by the
(1 + f(Za))TZ (which does strange things to
the notion of a homogeneous function). a
The intention is to regard f(Z ) as a twister function for a righthanded photon.
a
But, in accordance with (ii), such f(Z ) is really describing an
element of a sheaf cohomology group. appropriately cohomological. (say lP 'J1'+) fij
= fji
condition).
Thus we need to check that (A) is
Suppose we have a covering of the base space
by a number of open sets Ui.
Then on each Ui n Uj we need
such that on each ui n uj n uk, fik
=
fij + fjk (cocycle
If we use standard twister "coordinates" Za for the entire a
base space, then the fij will simply be homogeneous functions of Z , of
75
degree 4.
The
cocycle defining the righthanded photon can be taken to be just this collection of homogeneous functions. "coordinates"
i
a
We piece together our "bundle" by taking
for the portion lying above ui, where on each overlap (say,
the portion of the bundle lying above ui () uj) we have
~
a
=
{1 +
fij
(! e)} a . ~
~
We must first check the consistency of this with fij fji' Avoiding messy a e~a a e~a indices we have (A) : X = {1 + f (Z ) } Z and Z = {1 + g (X ) } X , and we need to check that g
= f.
Now
require (1 + g(Xe))(l + f(Ze))
Za
=
= 1;
{1
+ g(Xe)}~{l + f(Ze)}~Za, so we
that is
(because of the 4 homogeneity of g) so g(ZP) = f(ZP) as required. check the compatibility of (A) with fik = fij + fjk'
Next we
Again avoiding indices
we have, on the triple intersection region above ui () uj () uk relations
For consistency we require
p
i.e.
{1 +
p(Z ) }{1 + f(Zp)} 1 + f(ZP)
i.e.
q(ZP) = p(Z~) + f(ZP) cs required.
1
p
+ q(Z )
We need also to show that if the
cocycle is a coboundary, i.e., if fij = fjfi for each i,j, then the bundle
76
is the same as that for flat twistor space (fi being holomorphic throughout a Ui). For this we set Za = {1 + fi a)}!,. ( and find that the Z , so con
(i
structed, is defined globally over the whole bundle.
The compatibility over
each Ui n Uj follows by a calculation which is basically the same as the ones given above.
Finally we need the fact that if the bundle is the same as
that for flat twistor space (where here and above "the same" must be suitably interpreted in terms of "analytically extendible to"owing to the gaping holes in the fibres), then the {fij}cocycle is a coboundary. prove this, we simply reverse the above argument. a
existence of a global Z
To
The flatness implies the a
for the bundle.
The local Z for each patch (above i
Ui) must be related to Za by a formula za
=
{1
+ fi(z 6) }!r_Za from which i
i
follows (by the reverse of the above calculation) fij = fj  fi above ui n uj. The argument just given effectively shows that although (A) appears to be a highly nonlinear relation, the system of bundles so constructed has nevertheless a linear structure (given by simply adding the sheaf cohomology group elements).
This may be contrasted with Sparling's method of patching
together nonlinear gravitons:
For this, the cocycle condition fails.
This is a manifestation of the very
nonlinearity of the nonlinear graviton.
On the other hand, the
"4" in the
above construction is not essential for linearity and could be replaced by other powers; 1.e., · f o f h omogene i ty n, wi t h xa = {1 + f(Z 6)}l/nza.
When
n=l, "O", 1, 2, 3, ••• this is implicit in the SparlingWard construction. However, when n < 0 the global structure of the fibres would have to be something different (because (1 + z n ) 1/~ z > 1 as
Iz I
>co if n < 0, whereas
77
(1 + z n) l/n z  z as
Iz I
>
co
if n > 0).
In any case, the motivation from
"preservation of IDZ" exists only when n = 4. A direct construction of the righthanded Maxwell field
vAA' ~A'B' = O)
~A'B'
from the bundle structure has not yet emerged.
(satisfying
But the most
serious problem is that of fitting the righthanded and lefthanded ways of deforming '11'+ together into one bundle ••• ? is a method suggested by Hughston [12].
One possibility for doing this
However it is unclear, as yet, how
to extract left and righthanded Maxwell fields which do not interact with one another.
If this problem can be resolved, then among other things, it
might suggest an analogous approach to an ambidextrous graviton! 3.
REMARKS CONCERNING THE AMBIDEXTROUS PHOTON *
The information of a lefthanded Maxwell field can be stored [9] in a bolomorphic line bundle T over a suitable portion R of F'll', the Euler operator T being welldefined on T and pointing along the fibres of T.
Likewise, the
information of a righthanded Maxwell field can be stored in a !manifoldfibration T over R, but where now it is the holomorphic volume 4form which is welldefined on T instead of T. reference [9].)
~(=IDZ)
(See the previous section and also
In reference [12], Hughston suggested putting the right
handed construction on top of the lefthanded one, but it was not clear that this preserved any particular structure on T. operator I shall call DIV.
In fact it does, namely the
This is a map from vector fields V, pointing
along the fibration, to scalars, satisfying DIV(AV) A is a scalar field.
= ADIV(V) +
V(A), where
It differs from an ordinary div operation only in that
it need not act on any vector field which does not point along the fibration. Clearly, if
~
exists, then DIV exists, defined by DIV(V)
*Taken from Twister Newsletter 5 (July, 1977).
78
= div~
V, where
div~
on~
is the ordinary divergence based
3form V.J
~.
take the exterior derivative of V.J
to obtain a scalar). T x DIV(V)
(i.e.,
=
~.
use~
to convert V into a
and then divide out by
~
Also, if T exists, then DIV exists, defined by
[T,V] + 4V.
Let us now suppose that the 4dimensional space T, which is a fibration over R (all holomorphic), possesses just the structure DIV. sets {Ui} and lift these to give a covering {Ui} ofT.
Cover R by open
Define a
~.
Ti in each Ui which are compatible, in the sense above, with DIV.
l.
and an These
compatibility requirements can be expressed as and
4
(the latter being equivalent to
= 4).
div~_Ti
Now define fij
= log(~j/~i).
l.
We have£Tifij
=0
a 1cocycle on R. of the photon.
and fij  fik + fjk
=0
on Ui n Uj n Uk, so fij defines
This is the lefthanded (homogeneous of degree 0) "part"
Whenever this cohomology class vanishes,
Suppose, next, that
~
1cocycle 3form on R.
is global.
Define F ij
~
exists globally. We find
This is the righthanded (backhanded)(homogeneous
of degree 4) "part" of the photon. In fact, this is exactly what we had before (in Twistor Newsletter [10])
although phrased somewhat differently, so that we don't even notice
the( ••• )~ unpleasantness! i s to write the Z = (1 A
degree 4.
(Incidentally, another way of disguising this ~
+ 4g) Z as Z A
=
e
gT
Z, where g (Z) is homogeneous of
The power series do check!)
The trouble, of course, is the same as before: Part
3 and
(~
if there is a lefthanded
not global) then there seems to be no way of extracting an
79
H1 (R,0(4)) element to measure the nonglobalness of T (without bringing in more structure, that is).
This is the old problemthe righthanded part
appears to be charged with respect to the lefthanded part.
To decouple the
two parts, it seems that strict conformal invariance must be broken (e.g. Fij = (Tj  Ti) ...J~i or Fij = (Tj  Ti) ...J (~j  ~i) won't do because the cocycle condition fails).
Thus we must bring in I (the "infinity twister").
This works in principle, but is, as yet, inelegant. 4.
MASSLESS FIELDS AND SHEAF COHOMOLOGY*
In the first of these four Twister Newsletter articles I indicated that twister functions are really to be regarded as providing representative cocycles for sheaf cohomology, so the twistorial representation of a massless field of "helicity !!n is an element of H1 (X,0(n2)), where X is an 2
(open) region in lP n' (say lP n'+) swept out by the projective lines which correspond to the points of a suitable given (open) region Y in ¢M (say ¢M+) where the field is defined.
Woodhouse [7] showed how the spacetime field
could be obtained from the H1 ( ••• ) element, when the H1 ( ••• ) element is defined by means of Dolbeault [3,4] cohomology (i.e., "antiholomorphic" forms).
" A method due to Ward shows how to obtain the field from a Cech
cohomology cocycle fij using an extension of Sparling's "splitting" formula (cf. Ward [ 9]).
Ward's method essentially replaces the "branched contour
integral" described in the first section (cf. also Ward [10]).
Basically,
Ward's method is as follows: a
Let fij(Z) be homogeneous of degree n2 (with n ~ 0) and defined on a A A AA' nA'' ui n u.J (with {Ui} a covering of XC lP 'rr)' z = (w , nA,), w = ix etc.
Then nA'
...
a
nE,fij(Z ) (where n appears n+l times) is homogeneous of
*Taken (with minor modifications) from Twister Newsletter 5 (July, 1977). 80
degree 1 in n, so ("splitting") we get (Pi meaning "restriction," adopting a notational device due to Hughston, where square brackets denote skewsymmetrization):
whence
E'
4>jA' ••• D' (x,n)
11
1jljA' ••• D'E' Thus, due to the sheaf property
satisfies P[i 4>j ]A' ••• D' = 0.
4>jA' ••• D' (x,n)
and, since
~A'
P.~A' J ••• D'
••• D' is homogeneous of degree 0 in
must be a function of x only. follows.
Furthermore,
1l
(and global in
vAA' 4>A'B' ••• D' = 0
11),
it
readily
The case n < 0 is somewhat similar.
All this goes one way only:
from twistor function to spacetime field.
But using the method of exact homology sequences we can readily extend this to obtain (implicitly) a version of the "inverse twistor function," which goes from spacetime field to twistor function. SparlingPenrose method outlined in [1].)
(Compare also the Bramson
The following was inspired to
some extent also by Hughston [17], cf. also Lerner [18]. Let
~
be the (dual) primed spinvector bundle (excluding the section of
zero spin vectors) over ¢M.
a
Thus, a point of lF can be labelled (x ,nA,)
(with nA' ~ 0)except for points at infinity on eMwith xa complex. lF is a also a bundle over n {0}, the fibre over Z being the set of linear 2u
spaces containing Z •
81
lF
/~ U'  {O}
CM
Consider the following exact sequence of sheaves over lF: X 11
E'
>
n+l
zn'
o.
>
{B)
Here Z' stands for the sheaf of nindex symmetric spinor fields n
~A'
i
••• D'(x ,11Q,) which are holomorphic in x M'
freefield equations V
~A'B'
~A' ••• D'E' ~> ~A' ••• D' =
A'
=
o
11Q'
and satisfy the massless
a xA', 11A'
... D' = 0 (VM'
being merely a
ox
given~
11
E'
~A' •.• D'E' and we must check that is is onto,
••• we can always solve
for~
••••
One method is as A'
First, fix 11A'' and choose a primed spinor basis o
follows: n
and
' > Zn' is simply The map Zn+l
passenger).
i.e., that
i
A'
,1
A'
with
We have ~ 0' • • • 0' 0' = ~ 0 ' • • • 0' ' ljll '0' • • • 0' 0' = ~ 1 '0' •• • 0' ' • • • '
~l' ••• l'O' = ~l' ••• l'' but ~l' ••• l'l' is fixed only by the massless field
equations a a (  oxOl' ~1' ... 1'1' =aljll' XOO' ••• l'O'
a oxOO'
=~
1I
•••
1'
)
and a oxll' ~1' ... 1'1'
a ( oxlO' ~1' ... 1'0'
a a ~1' 1,), xlO' '' •
the integrability conditions for which come simply from the massless equations on
~A'
••• D''
choosing (say)
~l'
x 01 , = constant. 82
We can ensure a unique solution locally in ¢M by ••• l'l' = 0 on an initial 2surface x 11 , =constant, Second, allow
·u
A'
to vary locally in CC
2
 {O}; take
A'
1
to
depend holomorphically on n prescription for finding
A'
~A'
(
= oA' )
(say
1
A'
«
constant).
The above
••• D'E' ensures that it is holomorphic in nA,
though the dependence on nA' may be complicated owing to the fact that the functional dependence
of~
••• on the components xAO' is Rdependent.
(Incidentally, for simplicity, choose the unprimed basis constant.) To see what Tis we examine the kernel of the above map, i.e., find for which n
=0
(n
E'
~ 0).
We have (by standard lemmas)
= 0.
on~
••• become
Thus the dependence of f on x a is only through the quantity
AA' ix nA' (i.e., f is constant on twistor planes in eM), so
A
w
f
~A' ••• D'E'
nA, ••• nE,f, and the massless equations
ljiA' ••• E' nA' ~AA,f
E'
~·s
= f(w A,nA,) =
u
f(Z ) is a twistor function.
twistortype functions on lF; 'Ir 
T is the sheaf of
We see that
equivalently, of holomorphic functions on
{0}.
The short exact sequence (B) gives rise to the long exact homology sequence: ••• >
(C)
> I
I
Now, restrict Zn to be homogeneous of degree 0 in nA' (sheaf Zn(O));
' to be homogeneous of degree 1 in nA' (sheaf Zn+l(1)); correspondingly, Zn+l so that Twill be homogeneous of degree n2 in nA' (sheaf T(n2)), i.e., of degree n2 in Za • ab ove some Y C iiA' E
~2 
~M
{0}).
Choose Q C lF to correspond to the region of lF lying a
a
~
(so, for each (x ,nA,) E Q we have (x ,nA,) E Q whenever Then since H0 means "global sections" and since
Z~+l
is
homogeneous of degree 1 in nA' we have
83
' ' To study H1 {Q,Zn+l) we can use, for example, a "resolution" of Zn+l (where
dependence on nA' is irrelevant)
' 0 > 2n+l > FO,n+l > Fl,n > FO,n1 > 0 • this being exact, where Fp,q is the sheaf of fields
(D)
x~;::~F 1 (xi,nR,) with p
symmetric unprimed indices and q symmetric primed ones (freely holomorphic), so Fp,q is a coherent analytic sheaf, yielding Hr(S,F p,q ) = 0 whenever S is Stein and r > 0.
The maps are: ~A' ••• E' •
~ 1 XA'B' ••• E'
A
XA'B' ••• E'
I>
a
B' ••• E'
aA
B'C' ••• E' t> nC' ••• E'
0
B' A 'i/A a B'C' ••• E' •
and exactness is not hard to verify.
If n=O, the sequence terminates one
step sooner (and if n=1 we also have a short exact sequence
o > zo'
i >
Fo,o a> Fo.o
where 0 is the D'Alembertian). conformal weights 4.
of~
>
o
Note that (D) is conformally invariant, the
••• , x ••• , a ••• , n••• being, respectively, 1, 1, 3,
The sequence (D) (for which nA' is irrelevant) is of interest in the
study of massless fields quite independently of twister theory.
Here we
just use it to derive the fact that
whenever Y is Stein (as is the case if Y is the future (or past) tube in
¢M, or if Y is a suitable "thickening" of a portion of the real spacetime 8~
1
M·
(This follows because H (Q,FO,n+l(1)) then vanishes from the homo1
geneity degree 1 in nA', the fact that H (tlP 1 ,0(l)) = 0 and suitable general theoremsfor which thanks are due to R. 0. Wellsand because 0 I H (Q,FO,n+l(1)/Zn(1)) also vanishesmore obviously, because of 1 degree homogeneity.) The upshot of this is that since the first and last terms in (C) both vanish (Y Stein), we have (for n > 0)
which, because globality inn on the lefthandside implies constancy in n, almost establishes the required isomorphism between massless fields and
a1 •s
of twistor functions.
The remaining essential problem (the non
triviality and solution of which have been pointed out to me by M. Eastwood) is that all the sheaves refer to the space F, so far, and some subtleties are involved in projecting the cohomology groups down to
~.
The details
of this and other matters are left to a proposed later paper to be written jointly with M. Eastwood and R. 0. Wells, Jr.
In fact the argument can be
shown to work not only for the future tube (i.e., for X for suitable open regions in M.
= P ~+)
but also
(Another subtlety has been glossed over
in that the (x,n) description doesn't work for points at infinity.
But
this is unimportant because of conformal invariance and the fact that infinity can be transformed to somewhere safely inside the singular region of the field.) To deal with the cases n < 0 we need a different exact sequence:
a
0
> P_
~
a
2 ...!...> T ~ •• "a;;;C">
D' Z
m1
n
VDD'
=!>
zm > o.
(E)
85
Here Z
m
stands for the sheaf of unprimed massless holomorphic fields
i «~>A ••• o
Prn 2 stands for twistor functions
A
which are polynomials in w of degree at most m2 for each fixed nA'" fact that Z
m 1
>
Z
m
is onto, is basically an argument given if reference
[11]; the kernel is this: equations on which are
E' 1r
twistor functions
~A •••
c
the massless field
a
a aw
"w[DljJA] ••• c = 0, yielding ~A ••• c = ~·. u
as required.
The
a ·c f(Z) aw
The short exact sequence
0 > T/Pm_ 2 > Zml > Zm > 0 provides the long exact sequence 1 1 > H (Q,T/Pm_ 2 ) > H (Q,Zm1).
(F)
Choosing homogeneities 1, 0, m2, 1, respectively, and using the same argument as above, we get (Y Stein): 1
.
H (Q,T(m2)/Pm_ 2 (m2)). To deal with the extra complication of 0 > Pm_ 2 >
T > T/Pm_ 2
(G)
Pm 2 we use
> 0 to derive the exact sequence
(H)
(where k
= m2).
Now the sheaf Pk(k) vanishes unless k
~
0 (i.e., m ~ 2)
and, using a suitable Kunneth formula, we derive the fact (since homogeneity in nA' for each x is nonnegative) that we are concerned only with polynomial behavior in nA' for the two end terms in (H).
Consequently these
terms refer, in effect, to the "ordinary" cohomology of Q but with coefficients which are twistor polynomials Za ••• zYp
86
a ..• y
, homogeneous of
degree k, i.e., equivalently, symmetric ntwistors P . a ••• y
If Y describes
a spacetime region surrounding a "source tube," then we have
1 1 2 "6(k+l)(k+2)(k+3) = gm
, all of which have to
vanish if the H1 (Q,T) description is to work.
This corresponds to the
final map in (H) mapping to the zero element.
(Note that m=2 for the anti
a ••• y
selfdual Maxwell case, which is consistent with a statement made in the first section of this article.)
If Y is the forward tube t:M+, then the first
and last terms of (H) both vanish and the required isomorphism follows (but various details remain to be worked out). Alternative approaches ·to the cases n < 0 can be given in which a potential rather than the field is used. Pk and P
a ••• y
.
This avoids having to bring in
(Work by Ward, Hughston and Eastwood.)
Further work is in progress. Thanks are due to M. F. Atiyah, M. Eastwood, R. 0. Wells, Jr. and R. S. Ward. 5.
SOME REMARKS CONCERNING A GOOGLY* GRAVITON
The standard "nonlinear graviton" construction [8] provides a means of coding a general antiselfdual solution of Einstein's equations (or "rightflat spacetime") into the structure of a deformed twistor space. A general selfdual solution (or "leftflat spacetime") can be correspondingly coded into the structure of a deformed dual twistor space. However, it appears to be important, for the success of the twistor program as a whole, that the
~
twistor space be employable to encode the
~To those unfamiliar with the game of cricket, it should be explained that a googly" is a ball bowled in such a way that it spins in a righthanded sense about its direction of motion even though, to the batsman, it would appear that the bowling action would be such as to impart a lefthanded helicity.
87
information of both antiselfdual and selfdual types of vacuum curvature. The hope would be that by combining both parts of the spacetime (Weyl) curvature into the structure of
~
suitably deformed twistor space it might
be possible to represent, twistorially, the general (analytic?) solution of Einstein's vacuum equations.
1 shall here be concerned mainly with the
preliminary problem of attempting to encode the structure of a leftflat spacetime into a deformed twistor space. The essential difficulty in doing this appears to lie in the fact that the interpretation of a (complex) spacetime point as a compact holomorphic curve in deformed projective twistor space, with null spacetime separation being interpreted as intersection of the corresponding holomorphic curves, leads directly to the condition that many aplanes exist in the spacetime and that consequently the spacetime is rightconformally flat.
What seems to
be required is a different twistorial interpretation of a spacetime point. Now recall [13] the simple exact sequence that expresses the essential Poincare invariant structure of flat twistor space
~
where S is "unprimed" spinspace (the space of wA ) and "'* S is dual "primed" spinspace (the space of 'II' A'). The map i is wA o+ (wA ,0) and p is A a A (w ,'II'A,) ..... 'II'A" with Z = (w ,nA 1 ) . The standard twistorial representation ~,
of a (complex) spacetime point may be thought of as a crosssection of where ~ is regarded as a bundle over ~* with projection p, the crossa
section being lined up along the direction of the Euler operator T = Z and hence passing through the origin of
~
•
ataz G
Another way of viewing this
crosssection is as a map q from ~* to ~ such that the composition poq is the identity on 88
l*.
This is one method of "splitting" the exact sequence; there is a corresponding
~
method which is simply a map j from
property that the composition joi is the identity on $.
~
to S with the In the flat case,
the maps j and q contain the same information as each other, so they can each equivalently be used to represent a (complex) spacetime point. when
~
is suitably (mutilated
different.
However,
and) deformed, the information will be
The map q now gives the standard representation of a point in
the rightflat nonlinear graviton construction; and some corresponding "deformed" version of the map j ought to give a different difinition of a "point," the space of these "points" providing the required leftflat complex spacetime. There are many difficulties in making such an idea work in detail, however.
It seems that we need some noncompact version of Kodaira's theorem
[14,15] in ord~t to be sure that the family of deformed jmaps is actually 4dimensional.
We need some suitable (and yet unknown) regularity conditions
in the neighborhood of the image of i.
Indeed, it appears that this neigh
borhood contains the entire information of the selfdual curvature.
We
need, also, to be able to see why this neighborhood, if only infinitesimally deformed, is characterized by a 1st cohomology group of 0(6) twistor functions.
Work is in progress on these problems.
The idea, in order to proceed further, would be that with one and the same deformed twistor space we might have three alternative definitions of "spacetime point," the above two and a third which would be some kind of symmetrical combination of the other two, providing a general solution of Einstein's vacuum equations!
Assuming this solution to be asymptotically
"'
flat, the qdefinition would provide its Hspace and the pdefinition Hspace [16].
its
At the moment, however, this is all purely speculative.
89
References 1
R. Penrose,
"Twistor Theory, Its Aims and Achievements" in Quantum Gravity, eds: C. J. Isham, R. Penrose, D. W. Sciama, Oxford University Press, 1975,
2
R. Penrose,
"Twistors and Particles" in Quantum Theory and the Structures of Time and Space, eds: L. Castell, M, Drieschner, C. F. von Weiszacker, C. Hanser, Munich 1975.
3
J. Morrow and K. Kodaira,
Complex Manifolds; Holt, Rinehart and Winston, N.Y., 1971.
4
R. C. Gunning and H. Rossi,
Analytic Functions of Several Complex Variables, PrenticeHall, 1965.
5
R. 0. Wells, Jr.,
Differential Analysis on Complex Manifolds, PrenticeHall, 1973.
6
G. A. J. Sparling and R. Penrose,"The Twistor Quadrille" in Twistor Newsletter 1(4 March 1976).
7
N. M. J. Woodhouse,
"Twistor Cohomology Without Sheaves," in Twistor Newsletter 2(10 June 1976).
8
R. Penrose
Gen. Rel. Grav.
9
R. S. Ward,
"The Twisted Photon" in Twistor Newsletter 1(4 March 1976).
10
R. S. Ward,
"Zerorestmass Fields from Twistor Functions" in Twistor Newsletter 1 (4 March 1976).
11
R. Penrose,
Proc. Roy. Soc. A284, 159(1965).
12
L. P. Hughston,
"A Generalized RightHanded Photon Construction" in Twistor Newsletter 3 (7 December 1976).
13
R. Penrose and M. A. H. MacCallum,
14
K. Kodaira,
Am.
15
K. Kodaira,
Ann. Math. 75, 146(1962).
16
R. Hansen, E. T •. Newman, R. 'Penrose : and K. P, Tod, Proc. Roy. Soc. (to appear).
90
l•
31, 171(1976).
Phys. Reports 6C, 241(1973). J. Math. 85, 79(1963).
17
L. P. Hughston,
"The Twistor Cohomology of Local Hertz Potentials" in Twistor Newsletter 4 (1 April 1977).
18
D. Lerner,
"The 'Inverse Twistor Function' for Positive Frequency Fields" in Twistor Newsletter 5(11 July 1977).
R PENROSE Mathematical Institute 2429 St. Giles Oxford, England
91
R 0 Wells, Jr Cohomology and the Penrose transform
Penrose has described in various places his remarkable correspondence between Minkowski space and subsets of JP 3(¢)
first introduced in [5].
We will
recall briefly how the correspondence can be described in terms of a basic double fibration [12].
Consider complex flag manifolds defined by complex
subspaces of various dimensions of JP
=
:M
lF
{Sl c f/,4
dim sl
~
4
•
Namely, let
1},
{S2 c ~4
dim s 2 2}. 4 {Sl c s2 c a: : dim s 1
1, dim s 2
= 2}.
Then JP, :M, and lF are compact complex manifolds of dimensions 3, 4, and 5, respectively, with JP being a projective space, 1M a Grassmannian manifold and lF a more general flag manifold.
There is a natural double fibration of
holomorphic bundle mappings
(1)
inducing the Penrose correspondence t between F and 1M defined by t(p) = f3(a t
1
1
{p)) 1
{p) = Cl{f3
Thus points in
r
~
(p)}
F 2 (a:), ~
p E J
JP 1 (a:),
correspond to projective planes
correspond to projective lines in
92
p E F
r.
in :M and points in :M
we now introduce an Hermitian quadratic form ton ¢ 4 of signature (++). The pair (¢ 4 ,t) is called the space of twistors, and is denoted by '.n'. The compact complex manifolds above have homogeneous coordinates defined in terms of spinors and these coordinates are important for many calculations involving functiontheoretic and fieldtheoretic quantities on these manifolds. The conceptual results we want to describe will not normally involve these coordinates, but certain details of our proofs and the physical interpretation of some of our results certainly involve the twistor coordinates a great deal.
We now use the twistor metric t to define various subsets of
our three complex manifolds P,
lF,
and lM.
We will say that a subspace
s c ~ 4 is: a) positive if· t(v) is positive for any nonzero vector v e S, b) negative if t(v) is negative for any nonzero vector v E S, and c) null if
t(v) is zero for any vector v E S.
We now define
p+
{Sl c 1£4
dim sl
1, sl positive},
p
{Sl c f4
dim sl
1, sl negative},
p
{51 c f4
dim sl
1, sl null},
with JM+, nt:_, M, and lF+' lF_, and F defined in a similar manner.
Thus we
have induced diagrams:
~en closed """ compact
93
with the corresponding induced Penrose correspondences.
We find that M is
compactified real Minkowski space, and that SU(2,2), which acts transitively on the proper subsets F+' etc., defined above by of the conformal extension of the Poincare group. boundary component of both 1M+ and 1M and 1M is a complexification of M. submanifold of F
~.
induces on M the action
Moreover, M is a
which are bounded symmetric domains,
Slmilarly, F is a 6realdimensional
which is a CRsubmanifold of CRdimension 1, i.e. the
holomorphic tangent space ofF in F
is !dimensional (cf. [11]), and more
over, F is a boundary component of both F + and F _. P is a 5realdimensional hypersurface in P,
In the same manner,
and has CRdimension 2.
For
more details concerning these submanifolds and their construction see [12]. We remar.k that F,
F+'
F_,
1', .•. ,
all are complex manifolds, while P, F,
and M are CRmanifolds which means they have a certain amount of tangential complex structure inherited from their ambient complex manifolds P, F, and 1M. Our main problem is to describe solutions of various massless field equations on M in terms of intrinsic holomorphic data on F
F.
or subsets of
At the conclusion of [12], we indicated how holomorphic massless fields
of positive helicity on 1M+ could be described in terms of cohomology classes on P+' a result due toR. Penrose (cf. [7]).
In this paper we want
to indicate how this description can be extended to negative helicity and to weak
(hyperfunctio~
solutions on the real compactified Minkowski space M.
We will first recall the construction of [12] for positive helicity which necessitates introducing some terminology from the theory of complex manifolds (cf. [10]). Let X be a complex manifold and ~et
OX
functions on X, the structure sheaf of X.
be the sheaf of (local) holomorphiC If V
~
X is a holomorphic vector
bundle, then we will let OX(V) be the sheaf of (local) holomorphic sections
94
of V.
One sees easily that OX(V) is a locally free sheaf of modules over
the sheaf of rings OX, and that any locally free sheaf of modules over is the sheaf of holomorphic sections of a holomorphic vector bundle. f
consider a holomorphic mapping X+ Y of two complex manifolds. F
=
OX Now
If
Oy(V) is a locally free sheaf on Y, then we define the pullback sheaf
f*F to be OX(f*V), the sheaf of holomorphic sections of the pullback bundle.
On the other hand if
G is
any sheaf of abelian groups on X, then
we define a sequence of sheaves {f~G} on Y called the direct image sheaves under the mapping f.
th The q direct image sheaf of Gunder f, denoted by
f~G, is the sheaf generated by the presheaf
The stalk of f~ at P E Y is the direct limit (fqG)
*
= l!.m
p
Hq (fl (U), G),
U,:)p
which is essentially the cohomology along the fiber f 1 (p) (see [2), for a discussion of direct image sheaves).
We want to use these two basic con
structions of lifting and pushing down sheaves in conjunction with our double fibration (1) to construct certain fundamental sheaves on F,
rr,
and 1M. Let n
E ~.and
n2 let H +F be the hyperplane section bundle of
n2 raised to the power n2, i.e., the local sections of H are homogeneous functions in the homogeneous coordinates of F n2.
We define, for n E
of degree
~,
Op(Hn2),
a*ci>F (Hn2),
(2)
s.;a*Op (Hn 2 ) • 95
We note that we only use the first direct image sheaf in (2), since it turns out that (as described later in the paper)
Thus the sheaves SlF(n) and SlM(n) are the natural pullback and pushforward of the sheaf
sp (n)
in this context.
One can show that' i f n
~
0'
S.,.(n)
where ranka:Vn = rank 0 and where
0
n
n
2
(CE ) ,
th symmetric tensor product. denotes the n
Moreover, sections
of SlM(n) can be identified with hologorphic spinors of primed type on open subsets of lM, i.e., of the form
bA'B' ... D'} (cf. e.g. [5] and [12]). n
This is the basic local analytic data which will allow us to transform cohomology on subsets of P
to spinor fields on subsets of lM.
For n
~
0,
we will see that
and our construction breaks down.
We will consider the case n
~
0 later on
in this paper, and concentrate for the moment on the case n > 0 (which will correspond to positive helicity from the point of view of physics).
Thus
we have three basic sheaves on our three basic spaces, depending on the integer n.
On 1P we have "homogeneous functions" Sp(n) ; on F we have
"pullbacks of homogeneous functions" SlF(n), and on :M we have "spinors" SlM(n), canonically relat~d as sheaves by the geometry of the basic diagram (1).
96
We need one more ingredient, namely, the differential operators defined by the geometry of (1) which will correspond to certain field equations arising in classical mathematical physics.
Let T(F) be the tangent bundle
to F, and let Ta(1F) C T(1F) be the subbundle of vectors tangent to the fibers of the fibration a.
This induces a canonical surjection of the
dual bundles 1T
T* (1F) ~ T~(1F), and a differential operator
acting on differential forms of any degree.
Thus if we have smooth differ
ential forms of degree r on F, denoted by fr (lF),
then exterior differen
tiation can be considered as a mapping
of scalar rforms to T*(F)valued rforms. with
1T
a
Thus we obtain by composition
a differential operator
This differential operator corresponds to "differentiation along the fibers of a," and will play an important role below. introduce two new sheaves on F
s;
(n)
s;.
(n)
=OF
(a*Hn2
(OF
We want to use T*a = T*a (F)
and Jot respectively.
to
Namely,. consider
®(),
(a*Hn 2 ® T:),
97
i.e., the extension of our basic sheaves on F and 1M by the cotangent bundle along the fibers of
~.
This data allows us to construct the following
diagram, where all of the sheaves depend on the integer n, assumed positive for now:
(3)
The differential operator d on F
extends to bundlevalued differential forms
since the bundle a*Hn 2 has transition functions which are constant
along the fibers of a. with
a
a,
It extends to cohomology since da (anti)commutes
and we are representing the sheaf cohomology in terms of differen
tial forms via the Dolbeault isomorphism.
The vertical mappings I and Ia
are the canonical isomorphisms given by the Leray spectral sequence (cf. [2]),
which degenerates since
alsF
(n)
more detail later in the paper.
for j ~ 1.
= 0,
They are discussed in
Thus the differential operator d
induces
a
a firstorder differential operator V in (3) which turns out to be the a
zerorestmass operator acting on spinor fields (cf. [12] and [1]). mapping
P=
The
I o a* which transforms cohomology in P+ to spinor fields on
IM+ is called the Penrose transform. 0
One can show that p is onetoone and
that its image in H (IM+'~(n)) is the same as the kernel of va.
One thus
has the basic isomorphism for n > 0, 1 _P_ {holomorphic solutions of helicity s = n/2 of H (P+,FlP(n)) ~ the zerorestmass field equations on :M+}.
98
(4)
See [12] for a further discussion of this isomorphism.
The detailed
proofs will appear in ll]. The basic ingredients of the Penrose transform and the isomorphism (4) are contained in diagram (3).
We started with the fibration (1), and the We considered the sheaf Sl' (n) = 0p (Hn2 )
hyperplane section bundle H on l'.
which in turn generated canonically the corresponding sheaves SF(n) and S1M(n) on F
and ::M, where sections of S1M(n) turn out to be spinor fields
of spin n/2, for n > 0. bundle
T: along
The fibration F ~ l' gave rise to a cotangent
the fibers and a differential operator da along the fibers.
Representing cohomology in terms of differential forms then gave us the diagram (3), where a* is just the pullback of differential forms and d
a
exterior differentiation of differential forms along the fibers. is intuitively clear that if
= 0,
is
Then it
since a *41 will
Then general spectral sequence theory allows
us to conclude that the natural vertical mappings I and Ia (which correspond . explicitly to integration over the fibers of B), are isomorphisms. the induced differential operator V
a
1
E H (l'+,SlP(n)).
annihalates p
Thus
for
Intuitively, the pullback of the cohomology class when
integrated over the fibers of Bmust satisfy the field equations, and all solutions arise this way (see [12] for an explicit computation of the fact that VaP
111
Given a cohomology class
1
E H (F+,S1F),
defines a cohomology class in any neighborhood of
1
1
B (p), by restriction, and thus an element in (B*SF)p, sheaf S1M at p.
111
Loosely said, the cohomology class
111
the stalk of the
"restricted to the
fiber" gives an element in the "cohomology along the fiber," which is the desired evaluation mapping I.
Notice that the fibers of Bare simply
99
!dimensional projective spaces, and are, in fact, submanifolds of P. C p, for p E 1M (cf. (1)), and let 0 y
,. l(p) = y p
Let
be the sheaf of holop
morphic functions on Y •
Then one can show easily that the basic sheaf
p
S1M(n) is the sheaf of holomorphic sections of the holomorphic vector bundle
V
n
~
1M, where the fibers of the bundle V for p E 1M is given by n,p
vn,p
{5)
where Hn2 means Hn2l y
which is the same as the hyperplane section p
bundle of P 1
vn
~
Yp' raised to the (n2) power.
We could have defined (6)
=
but it is not as clear from this definition that Vn is a welldefined smooth or holomorphic vector bundle as p varies in for p
~
q.
~.
since Yp might intersect Yq
That is the reason for using the space F
fibers apart)
and the direct image sheaves.
simply from this point of view: 4» to the submanifold Yp C P+'
(which splits the
But the Penrose transform says
1 n2 take 4» E H (P+' Op (H }} , and restrict
for p E lM+' thus obtaining an element in
H1 (Yp,OY (Hn2 )), using, for instance, differential forms and their p
restrictions to submanifolds as differential forms to represent cohomology and the restriction mapping on cohomology.
Here is why we assume n
since in that case Hl(Yp,OY (Hn2)) "H1 (P 1 ,0p (Hn2 )) p
1
"'H0 (P 1 , 0(Hn)) ~(!5 n(a:2).
Whereas if n < 0, we have
100
~
0,
1
0 H.(J.,YQI p p
n2
)) .. 0
1
by e.g. the Kodaira vanishing theorem (cf. [10)), i.e. 6*(SF (n))
0, i f
n < 0, as indicated earlier. So consider the problem of evaluating a cohomology class and obtaining a spinor field for n < 0.
~
1
E H (F+,SF(n))
Simple restriction to the fibers Y
p
by the above process gives zero, as the vector space where the evaluation takes place itself vanishes!
Nevertheless Penrose has given contour
integral formulas [7] which show how to evaluate explicit representations of such cohomology classes to obtain fields of the form (using twistor notation)
~AB. • .D(x) =
f
a a A E' awA•. • • awD f(Z )n dnE',
(7)
where f has homogeneity n2, and n < 0, and the integral is over an appropriate c~r.tour in , 1 (x)
= YX •
We want to evaluate the cohomology class
~
considered above in terms of
the leading term of an expansion of the cohomology class in a power series expansion about the fiber Y
p
T
1
(p), for some p E ::M+.
and let p remain fixed in this discussion. sections of
OF
We will let Y = YP
Let Iy be the ideal sheaf of the
which vanish on Y, and consider the short exact sequences of
sheaves:
(8)
Now Im/Im+l can be identified with the ~ symmetric power of the dual to y y 101
the normal bundle to the embedding Y C F
(cf. Grauert [3]).
that if N + Y is the normal bundle to Y in F, the hyperplane section bundle of F thus N*
~ Hl (f) H 1 •
We can see
then N ~ H ~ H, where H is
restricted toY (cf. Penrose [6]) and
It follows that
and
Now tensor SI'(n)
(as OF modules) with the above sequences (8) obtaining
0 + Iy ®SF (n) +SF (n) + Oy(H 0 +
I~
n2
) + 0,
® SF (n) + Iy ® SF (n) + ' \ (Hn 2 (8) N*) + 0,
(9)
Look at the associated long exact sequences of cohomology on F+: 1 io 1 ro 1 2 + H (F+' Iy®SF (n))+ H (F+' SF (n))+ H (Y,Oy(Hn )) + 1 2 il + H (F+, [email protected] SF (n)) 
1 H (F+' Iy r
4 + Hl(F
I m+l +' y ®SF (n))
im
=+
®SF (n)) +
H1 (Y ,Oy(Hn 2 ® N*)) + •.• ,
1 m' H (F+' Iy ®SF (n)) +
... , (10)
Now we note that Y ~ lP 1 (a:)
and if we now assume that n < 0, then by
Kodaira's vanishing theorem
for m that
<
n.
* N ~
Recalling that Y 1
H
vn,p
= Yp,
for p E IM+ in the above arguments and
1
<:9 H , we find that if we let, for n
< 0,
Hl(Yp' ~ (Hn2 ®0n(H1 $ H1)))'
(11)
p
then we find that
vn,p Thus the vector space V is a vector space of spinor type, for each n,p (although it is clear that to define V by (11) we can have p E lM, n,p
p E lM+
By defining Vn' for n < 0, by (11) to be the "holomorphic vector bundle with fiber V , " we see that we can evaluate a cohomology class n,p 1 for each cp e H (lP+,slP (n)), for n < 0, on Y to give an element of v p n,p p E IM+.
s~e
We will
later that Vn so defined is indeed a bundle, but we
will discuss the evaluation procedure first.
1 I f cp E H (lP+' SlP (n)) is
given, for some n < 0, then we see that in (10), r 0 (cp) with cp 1 E H1 (lP+' !YP ®SlP (n)). evaluated cp on YP.
= 0,
so cp
=
i 0 (cp 1 ),
I f n = 1, then r 1 (cp 1 ) E Vl,p' and we have
If n < 1, then r 1 (cp 1 )
= 0,
and cp 1
=
i 1 (cp 2 ), and we
Finally, by induction, we arrive at r
n
(cp
n
) E V • n,p
One checks that the element of V obtained in this manner is independent n,p of the choices made, and that a canonical evaluation of cp on Y has been p
made.
This corresponds to the (n)th term of the Taylor expansion of cp
about the fiber Y , and is the leading term in such an expansion p
103
(cf. Grauert [3]).
It depends on the normal bundle information up to the
(n)th order, just as does the formula of Penrose [7], but it does so in an intrinsic and equivariant manner. stand why
U pElM
lM.
Vn,p , for n
<
The only remaining problem is to under
0, is indeed a holomorphic vector bundle over
To do this we need to carry out the above evaluation process, and hence
the idealtheoretic arguments in a uniform and smooth manner, not simply one fiber at a time.
The basic idea is to consider
and
defined by
y
=
u
(Y
pE:M+
x {p}),
p
a complex submanifold of complex codimension 2 of P+ x :M+ with the property One sees readily that Y C P+ with the property that Nly
*
+ Y p
x :M+
has a normal bundle N+ Y
is the normal bundle of Y in P. p
p
In fact
N is given by 1r (H E9 H)
ly•
projection, and H + P
is the hyperplane section bundle of P, as before.
where 1r:P+
x :M++ P+
We now consider 1r * (Sp(n)) on P+ x :M+' ideal sheaf of Y in P+ x lM+. Sp(n) by
'If
* (Sp(n))
is the natural product
and let Iy c
op
+
X
:M be the
+
In (9) and (10) we replace Iy by Iy and
obtaining analogous short and long exact sequences on
P+ x lM+.
It follows from the same. kind of vanishing theorem considerations
on :M+
that we have the following diagram:
10/,i
x P+
1 H (l'+
X
(12)
Here Tis the Taylor expansion evaluation using the analogues to (9) to (10), and where N0* = H1 E9 H1 is a bundle over lP, whose pullback N* = the conormal bundle on
Y.
11
*N* is 0
The top horizontal isomorphism comes from a
comparison of the pullbacks by a *
and
11
* respectively, the mapping C is the
natural contraction obtained by tensoring H with H1 , and powers of same, and I is our uirect image map (integration over the fiber again).
The
resulting mapping p we call the Penrose transform for negative helicity. (i.e. for n < 0).
If we define Vn• for n < 0, to be the holomorphic vector
bundle on 1M such that
then we see that (13)
is a welldefined mapping and that
vn,p as desired.
Moreover, the sections of Vn can be identified with spinor
105
fields of unprimed type, i;e, spinors fields of the form
{~AB ••• D}
on open
n
subsets of lM. This is thus the Penrose transform p extended to negative helicity. must extend the differential operator d
a
One
to this situation obtaining mappings
from the various spaces in (12) and finally that the image of p satisfies the zerorestmass field equations. carried out here (cf. [1]).
This is not difficult, but will not be
Note also that we have neglected the case n
= 0.
This is a borderline case, and the Penrose transform p is welldefined with th only the zero order Taylor expansion, but the image of p satisfies a second order differential equation (the scalar wave equation), but this is not readily apparent from the above analysis.
This case is considered in
detail in [ 1]. Finally we want to indicate how one can extend the Penrose transform to obtain solutions of the field equations on M which are not smooth. result is that hyperfunction cohomology on P C P
The basic
is transformed onto all
hyperfunction solutions of the zerorestmass field equations on M C IM.
To
do this we need to use relative cohomology and understand something about taking boundary values of cohomology classes defined in P ± on the common boundary P.
First we want to see how we can obtain realanalytic solutions
of the field equations on M.
We will consider the case only where the
helicity is positive, since the analysis is simplest there, as we see from the previous discussions. Let
~
PU "" a(B
106
1
be an open neighborhood of M in 1M , and let 1FU (lfu)).
Thus we have our basic diagram
with our basic sheaves SlP(n), SlF(n),
and SlM(n) defined on these open sets.
For simplicity we will omit the notational dependence on n, assumed positive. We have the Penrose transform defined by
Now if U is a fundamental system of neighborhoods of M we see that p expends to the direct limit
p:
l!:m H1 (lPU, SlP) lPU:>p
+
l!:m H0 (~, SlM).
(14)
~=>M
We see that the direct limit on the right is, by definition, realanalytic spinor fields on M.
It is not clear from this analysis that the real
analytic solutions of the massless field equations are precisely the image of p in (14) , but this is nevertheless the case (see [ 13]).
Let us denote
(15)
107
the subscript
A denoting
realanalytic.
Thus we have p defined on real
analytic data on P. If S C X is a closed subset of a topological space X, and f is a sheaf of abelian groups on X, then we will let H~(X,F) denote the relative cohomo~
of X with coefficients in
F relative to the subset X.
often referred to as cohomology with .supports in S.
This is also
There is a long exact
sequence of relative cohomology
which shows that the relative cohomology groups are the obstructions that a cohomology class on X  S be the restriction of a cohomology class defined on all of X (cf. Schapira [9], for instance, for a discussion of this as well as of hyperfunctions, which we will use below.)
Hyperfunctions on M
can be defined as
and it is a deep result of Sato that this is the same as the dual of the realanalytic functions A(M)', where A(M) is equipped with the inductive limit topology of Frechet spaces of holomorphic functions, A(M) "' 1~
0(1Mu).
lMu.:::>M
Knowing that B(M)
= A(M)',
it is then easy to see that B(M) contains in a
natural manner the distributions on M, V'(M), since A(M) is dense in V(M). But the interpretation of hyperfunctions as relative cohomology gives a powerful means of working with hyperfunctions (cf. Schapira [9]).
We will
use this representation of hyperfunctions to indicate how the Penrose transform can be extended to the level of hyperfunctions (cf. the extension 108
of the Fourier transform from smooth functions to distributions of the appropriate type). on P.
First we need some appropriate "hyperfunction cohomology"
We consider +
1
and one knows that H ( lP, S lP)
••• '
0, and thus
2
Intuitively, Hp(lP, SlP) corresponds to the "jump" of the "boundary values" 1
1
of H (lP+' SlP) and H (lP_, Sp) on P, and this can be given a precise interpretation.
Ordinary hyperfunctions on ~n can also be considered as jumps of
boundary values of holomorphic functions defined in open subsets of ¢n  ~n (cf. Schapira [9]), and we have the same phenomenon at the level of the first cohomology groups.
2
In particular, one can interpret Hp(lP, Sp) as
intrinsic hyperfunction cohomology on p determined by the apcomplex on p acting on hyperfunctionvalued (0,2)forms.
See [8] for a discussion of
these _points. If Y C X is a real submanifold of a complex manifold X of real codimension r, then [Y] is the current of integration over Y.
It is a current (differ
ential form with distribution coefficients) of degree r (cf. Harvey [4]). Since X is a complex manifold, any current ' of degree r can be written uniquely as the sum of currents of specified bidegree or type
just as is the case for differential forms with smooth coefficients.
So in
Particular, if Y is the submanifold above of codimension r, then
109
[y]
[y]r,O + ... + [y]O,r ,
and [Y]O,r is a singular current of type (O,r) with support on Y. We now apply the above analysis to our submanifolds P c lP,
etc. , to
obtain an embedding of realanalytic cohomology into hyperfunction cohomology as follows:
(16)
0 1
We see that, for instance in the first case,~A[P] •
is indeed a aclosed
(0,2)current on P with support in P, and will represent a class in 2
!ip(lP, SlP) , and similarly for the other cases.
Here we are representing
~(lP, SlP), as aclosed (0,2)hyperfunction forms on lP with coefficients in SlP' modulo aexact forms of the same type. usual Dolbeault isomorphism.
This is a generalization of the
It would be impossible to represent relative
cohomology (or cohomology with support in P) in terms of smooth differential forms.
A hyperfunction
form here is a generalization of a current, allowing
hyperfunction coefficients instead of distribution coefficients, so a current 0 1
such as [P] '
is certainly a hyperfunction form of type (0,1).
We will have extended thP. Penrose transform p to the hyperfunction level if we can extend the mappings a * and I in (15) to the right hand column in 110
(16) to be compatible with the horizontal injections in (16).
This is
possible, and the principal tool is to introduce a "current of integration to be denoted by [CL] 0 • 3 , which is defined in the
along the fiber of
CL"
following manner.
Consider the pair of fibrations:
p
+lP
the left hand fibration F
~ P has fiber s 1 , and the right hand fibration
has fiber lP 2 (0:) , and we denote the inclusion of fibers by FP C lF P. for pEP.
Thus F
p
is a 3codimensional submanifold of lF
a current of integration on lF [F ] 0 • 3 • p
p
p
and as such defines
denoted by [F ], which has a component p
This is a family of currents parametrized by p E F.
We denote
0 3
this Fparameter family of currents symbolically by [~ • , and one shows that one can form the tensor product of this measure along the fibers with 2 currents or hyperfunction forms representing elements of Hp(lP, SlP)
define hyperfunction forms on lF of type (0,5).
to
This is similar to
constructing product measures on Cartesian products of measure spaces. 2 is the required "pullback" of the cohomology class in HP (lP, SlP) 5
HP ( lF, SlF) , and we denote this mapping by
~*
CL
This
to
One can also extend the Leray
spectral sequence argument and direct image mapping I to the case of cohomo~
logy with supports in F and M, and extend the mapping I to a mapping I giving us then the following commutative diagram:
111
and we have defined p
= "'loa"'* •
It is again necessary to extend the differen
tial operator dato this setting which is not difficult, and one obtains the result that
maps holomorphic hyperfunction data on P to hyperfunction solutions of the field equations on M (in positive helicity in the case above), and moreover, all hyperfunction solutions on M arise in this manner (see [13] for the proof
of this theorem).
There are numerous problems which remain in this area, e.g. extending the Penrose transform to coupled interactions of zerorestmass fields and YangMills fields, or to the more general geometric situation involving deformations of subsets of P which arises in Penrose's nonlinear graviton [6]. There are various versions of inverse transforms to the Penrose transform, which have been developed by Hughston, Penrose, and Ward (cf. various issues of the Twistor Newsletter from Oxford), and such inverses will be discussed in more detail in [1].
We will close with one philosophical remark.
The
power of the Fourier transform is that it converts problems concerning differential operators to problems involving algebraic operators. Penrose transform has this feature also.
112
The
Namely the differential operators
of mathematical physics get replaced by the algebraic operators involving sheaf cohomology.
It would be worthwhile to understand this phenomenon
independent of the particular differential equations that one wants to study, as has been the case here, and to see how close an analogue of the powerful Fourier transform this new transform of Penrose's really is. References 1
M. Eastwood, R. Penrose, and R. 0. Wells, Jr., "Cohomology and massless fields," (to appear).
2
Roger Godement,
Topologie Algebrique et Theorie des Faisceau, Hermann & Cie, Paris, 1964.
3
H. Grauert,
"Uber Modifikationen and exzeptionelle analytische Mengen" Math. Ann. 146(1962), 331368.
4
R. Harvey,
"Holomorphic chains and their boundaries," Proc. Symp. Pure Math. Vol. 30, Part 1, Amer. Math. Soc., Providence, R.I., 1977, 309382.
5
R. Penrose,
"Twistor algebra," J. Math. Phys., 8(1967), 345366.
6
R. Penrose,
"Nonlinear gravitons and curved twistor theory," General Relativity and Gravitation, 7(1976), 3152.
7
R. Penrose,
"The twistor program" Rep, on Math. l'hys. 12", 65(1977).
8
J. Polking and R. 0. Wells, Jr.,
"Boundary values of Dolbeault cohomology classes and generalized BochnerHartogs Theorem," Abhand. Math. Sem. Univ. Hamburg, 47(1978), 124.
9
P. Schapira,
Theorie des Hyperfunctions, Lecture Notes in Math., Vol. 126, SpringerVerlag, BerlinHeidelbergNew York, 1970.
R. 0. Wells, Jr.,
Differential Analysis on Complex Manifolds, PrenticeHall, Inc. Englewoolt.Cliffs, N.J., 1973.
10
113
11
R. 0. Wells·, Jr.,
"Function theory on differentiable submanifolds," Contributions to Analysis, Academic Press, 1974, 407441.
12
R. 0. Wells, Jr.,
"Complex manifolds and mathematical physics," Bull. Amer. Math. Soc. (to appear).
13
R. 0. Wells, Jr.,
"Hyperfunction solutions of the zerorestmass field equations," (to appear).
Acknowledgment The research is supported by NSF MCS 7803571.
R 0 WELLS, JR Department of Mathematics Rice University Houston, TX 77001
L P Hughston Some new contour integral formulae
1.
INTRODUCTION
The natural setting(l) for twistor theory and in particular, those aspects of the theory concerned primarily with zero rest mass phenomena  is complex 3
projective threespace P •
In connection with the problem of understanding
the role of massive fields in the theory it is of great interest that much of the analysis generalizes, in a remarkably straightforward way, when one goes up from P 3 to PZn+l, where n is any positive integer.
The basic setup
is as follows. a n+l Let S denote C regarded as a complex vector space; and denote by S , a
a'
S
, and S , a
the dual space, the complex conjugate space, and
r~spectively
the dual complex conjugate space to Sa.
By a generalized twistor we mean a
point in the space (Sa,Sa,), i.e. a pair (wa,~a,) with wa £Sa and ~a' £Sa'" Let us denote the space (S twistor space.
2n+l
The dual space to Ta Z
=
(w
a
.~
,s a ,)
a a by T ; in the event that n = 1, T is clearly
In the general case we have Ta: c2n+2 , and the associated
projective space in P
a
a
• For any generalized twistor

a
,) we define its complex conjugate by Z
a

<~
a
a' ,w ) •
Clearly, we
a' have Z £ T. If W = (a ,T ) is a dual twistor then the inner product a a a a a.._ a a' In particular, the norm of Za is Z wa is defined to be w cra + ~a,T
defined to be
zaz.
This norm has signature (n+l,n+l).
We shall denote by
a + a T ,T , and N those portions of T for which the norm is positive, negative,
and zero, respectively; the associated regions of the projective space PT will be denoted PT+, PT, and PN, respectively.
115
The Grassmann variety of projective nplanes in P In the case n
=1
2n+l
. 2 has d1mension (n+l) ,
the Grassmann variety can be regarded  and this fact is
perhaps the true starting point for modern applications of algebraic geometry to problems in theoretical physics  as complex Minkowski space, with some structure added in at infinity( 2).
A similar interpretation is valid for
general n, and the "finite" points of the Grassmannian in that case can be a represented by points in the complex vector. space S
being denoted x determined by x
aa' aa'
A
twistor
Z
a
= (wa ,na ,) a
if and only if w
~
a' S , a typical point
lies on the projective nplane
ixaa ' TI
a'.
Let us denote the abovementioned Grassmannian by M.
It is noteworthy
that certain categories of field equations can be introduced on M which  in an entirely natural and reasonable way  generalize analogous field equations on Minkowski space.
Suppose one considers, for example, the neutrino
equation
VAB'
B'
(1.1)
0
where cpB' (xAA') is the neutrino wave function.
How can equation (1.1) be
suitably generalized to arbitrary values of n?
First, we rewrite (1.1) in
the form VA[A'B']
0.
Then, we generalize by making the following sub
stitutions:
+
va a '
where vaa'
B' (x
AA'
)
(1.2)
When these substitutions are made in (1.1) the result
is va[a•¢b•] = 0 which one can tentatively regard as the proper generalization of (1.1).
116
(1.3) It
should be noted moreover that equation (1.1) implies automatically the wave equation D~B'
0.
VA[A'VB']B~C'
0, and thus generalizes, using (1.2), to:
The wave equation can be written in the form
(1.4) For n
~
1 equation (1.3) likewise gives equation (1.4), and therefore for
reasons to be justified in Section 3 we can regard (1.3) and (1.4) together as comprising the correct generalization of the neutrino equation. In what follows certain aspects of the geometry of higher dimensional twistortype spaces will be discussed, and in particular some new contour integral formulae will be introduced  modelled after the extraordinary twistor contour integral formulae that Penrose used [3] for solving the zero rest mass equations in Minkowski space  in order to solve the combined system of equations (1.3) and (1.4), and also to treat certain related systems of equations.
2.
A GENERALIZATION OF THE TWISTOR EQUATION
From a Minkowski space viewpoint projective twistors can be characterized as solutions of the differential equation this equation( 3 ) is
vA'(A~B) = 0. Indeed, the solution of
~A= wA ixAA'~A'' and the associated twistor is A.~A,).
determined by the fixed spinor pair (w
Now consider the equation (2.1)
It is not difficult to check that in the case n = 1 equation (2.1) reduces to the twistor equation.
The following result establishes the connection
between solutions of (2.1) and the generalized twistors introduced in the Previous section: 117
The generat eotution of equation (2.1) ie given by
Theorem 1
aa' ix
1.41here
wa
Proof
and
'IT ,
a
'ITa'
(2 .2)
'
are aonetant ( 4 ).
Differentiating (2.1) one has
(2. 3)
which, upon interchange of the index clumps aa'and bb', can be written as
(2.4) Since Vaa' and Vbb' commute, the lefthand sides of equations (2.3) and (2.4) are equal; therefore:
(2.5) Transvecting equation (2.5) with ob gives c (2.6)
On the other hand, transvecting (2.5) with 'Vba'
'Vdb'~
d
which, substituting b
~a,
o~,
one obtains (n+l)'Vbb'
'Vda'~d
then reads:
(2.7) Taken together, equations (2.6) and (2."1) imply
0 '
showing that .
l.n X
aa'
(2.8)
'Vdb'~d is constant; it follows, from (2.1), that ~a is linear
Inserting the most general linear expression for ~a into (2.1), a
short calculation verifies that (2.2) is, as desired, the general solution. 118
D
It should be observed that points in the dual space T
a
can be represented 
up to proportionality  by solutions of the equation (2.9) which is the primed analogue of (2.1); following the pattern of Theorem 1 it can be shown that the general solution of (2.9) is of the form
n
a'
T
a'
+ ix
where the pair (a ,T
a'a
a'
a
3.
a
(2 .10)
a
) corresponds to a point
W in the dual space TN. a .....
CONTOUR INTEGRAL FORMULAE
Let us denoteby px the operation of restriction to the projective nplane in
PT corresponding to the point x
aa'
in M.
Thus for any twistor Za ~ (wa,~ ,) a
we write: (ix
aa'
~
a
.. ~ ,)
(3 .1)
a
Similarly, if f(Za) is a function of Za, then we shall write p f(Za) X
f(p Za). X
Now consider the contour integral expression(S) (3.2)
where $(Za) is a holomorphic function  the singularity structure of which will be indicated, by means of some examples, in the next section  taken to be homogeneous of degree n2 in za. ~~
The differential form ~~ is defined by .(3.3)
Note that~~ is homogeneous of degree n+l in ~a'' and thus that the quantity
119
appearing under the integral sign in (3.2) is  taken in its entiretyhomogeneous of degree zero. Theorem 2
The
fie~ ~a,(x)
defined in formula (3.2) satisfies equations
(1.3) and (1.4). Proof
Using the differential chain rule it is straightforward to establish
the relations
(3.4) where nb
= a/awb,
the minus sign being introduced for later convenience.
Equations (3.2) and (3.4) together readily imply the field equations (1.3) and (1.4).
0
More generally, one can consider holomorphic functions homogeneous of degree n1r, where r is any integer.
When r > 1, a formula similar to
(3.2) applies, only the number of ncoefficients is r, and the resulting symmetric field
For r
~a'b'
••• c'(x), which has r indices, satisfies:
vd[d' ~a']b' ••• c'
0
vd[d' ve']e ~a'b' ••• c'
0
= 0,
(3.5)
no ncoefficients appear, and the resultant scalar field
~(x)
~t~f~s:
(3.6) Finally, for r < 0 we find solutions of the equations
120
vd'[d ~a]b ••• c
0
vd[d' ve']e ~ab ••• c
0
(3.7)
where the symmetric field ~
=
b
a .•• c
the operator n
4.
~
b
a ••• c
f "'"' "'
Px 1f a 1fb ••• TI c W(Z
is defined by a.
)~TI
(3.8)
appearing r times in the integrand.
a
EXAMPLES
We shall conclude with an illustration of the procedures outlined above.
It
is straightforward to construct, for each value of n, fields which generalize the notion of an given by n
=
elementa~y state( 6 ). Let us first take the simplest case,
2 and r
= 0.
For our twister function we take (4 .1)
where Pa.' QS' and Ry denote fixed points in the dual space.
For their
"spinor parts" we shall write p
a.
(4.2)
and thus, using definition (3.1), we can put a' px p a.Za.= p na' b' Px Qeze= q 1fb I c' Px Ry zY= r TIc' where p a'
•
b' q
and r
p
a'
b' q r c'
c'
I
p a' + ixa ap a b' ixb'bQ Q + b c' ixc ' cR R + c
)
are solutions of equation (2.9).
(4.3)
Equations (4.1)
and (4. 3) together give: (4.4)
121
Inserting this expression for p
X
~(Za) into the contour integral formula (4. 5)
we obtain, upon performing the integral, the result ~(x)
with 1 that
=
= 1[£ a 'b' c ,
(2~i)
~(x),
2
p
a'
b'
q
r
c' 1
(4.6)
]
, the contour being an S
1
1
x S •
It is not difficult to verify
as given in (4.6), does indeed satisfy equation (3.6).
Further examples are readily constructed: for instance, using the relations ()p
X
~(Z
a
)/Clp
a'
= ~a ,[p a' ~a ,] 2
[q
b'
~b'
r
c'
~ ,]
1
c
(4.7) which follow at once from equations (4.3) and (4.4), one sees that the twistor function (4 .8)
generates according to formula (3.2)  the field ~a' (x) =
1£ a 'b' c , q
b' c' r
[E a 'b' c , p
a' b' c' 2 q r ] ,
(4.9)
and that this field satisfies equations (1.3) and (1.4). Generalizations of these examples for other values of r, and for higher value of n in particular, to cases where n is odd, for which (as will be discussed elsewhere) a spacetime interpretation is readily available for the resulting formulae  are not difficult to construct.
122
1.
See, for example, reference [1], section VI.
2.
The Grassmann (or Klein) representation for the system of lines in p 3 as a quadric hypersurface in P 5 has been described [2,p.390] as an
"
illustration of the use of geometry ••• that has proved extraord
inarily useful and suggestive, as well as being capable of very wide application".
These words, written in 1952, are no less true today,
especially when one considers the remarkable position occupied by this representation within the framework of twister theory. 3.
The twister equation is discussed in considerable detail in reference [1], section V.
4.
See also references [3], [4], and [5].
It is perhaps worth mentioning, as a simple corollary to TheoPem 1, that a necessary and sufficient condition for a field ~a(x) to satisfy equation (2.1) is that it should satisfy
for all values of cra. n
Another point worth noting is the following.
For
= 1 the twister equation 'VA 1 (A~ B) = 0 is, in fact, a special case of
a more general differential equation  of wide interest, especially when it is considered within the context of general relativity  known as the
geodesic sheaPjPee condition: ~A~B 'VA'A~B
= O.
However, when n > 1 the
connection between the twister equation and the geodesic shearfree condition is lost, the appropriate generalization of the latter being:
Which does indeed, for n
= 1,
reduce to the geodesic shearfree condition.
Solutions of the equation above can be generated (as will be described
123
elsewhere) through the consideration of the intersections of projective · p 2n+l w1t · h certa i n types o f comp 1 ex ana 1yt1c · var1et · i es 1n · np 1 anes 1n P2n+l. 5.
Twistor contour integral formulae alluded to very briefly in Penrose's epoch making 1967 paper [l,p.347]  are first introduced and discussed in detail in reference [3], a magnificent opus from which spring ultimately many of the ideas which have dominated the outlook of this conference and which are destined by their very nature  rooted as they are ir. the fertile soil of algebraic geometry and complex analysis  to play an increasingly significant role in theoretical physics for many years to come.
6.
For an account of the notion of elementary state, see Penrose and MacCallum [4,p.279].
Elementary
s~ates
are the prototypes of non
singular, positive frequency, asymptotically wellbehaved, normalizable wave functions; and they emerge in a strikingly natural way within the context of twistor theory.
ACKNOWLEDGEMENTS
I would like to thank R. Penrose, K.P. Tod, R.S. Ward,
N.M.J. Woodhouse and other colleagues at The Mathematical Institute, Oxford, for useful conversations in connection with the material described herein. This work has been supported by the Science Research Council, and by a Junior Research Fellowship at Wolfson College, Oxford.
REFERENCES [1]
R. Penrose,
[2]
J.G. Semple and G.T. Kneebone,
Twistor Algebra, J.Math.Phys.
345 (1967).
Algebraic Projective Geometry, Clarendon
Press, Oxford (1952). 12~
~.
[3]
R. Penrose,
Twistor quantisation and Curved SpaceTime, Int. J. Theor. Phys,
[4]
!, 61 (1968).
R. Penrose and M.A.H. MacCallum, Twistor Theory: An Approach to the Quantisation of Fields and SpaceTime, Phys. Reports, 6C, 241 (1972).
[5]
R. Penrose,
Twistor Theory: Its Aims and Achievements, in quantum Gravity: An Oxford Symposium, C.J.Isham, R. Penrose, and D.W. Sciama editors, Clarendon Press, Oxford (1975).
L.P. Hughston The Mathematical Institute Oxford England
125
C M Patton Zero rest mass fields and the Bargmann complex structure My purpose here is to explain certain aspects o! the twistor description of zerorestmass fields from the point of view of a different complex structure on twistor space.
In particular:
=
(a)
the "2" in "n2" where n
2s, s the helicity.
(b)
normalizable fields in terms of holomorphic functions on all of twistor space.
(c)
the relative unnaturality of unprimed spin fields.
(d)
the inner product on cohomology.
We will begin by considering an inneL product structure on holomorphic and antiholomorphic functions on
~
2
, when ¢
hermitian scalar product, denoted <,
>.
2
is equipped with a positive definite Consider the space
where d~ is the normalized Euclidean measure on ~ 2 , i.e., if
Fn
E& F where n=O n
= {f homogeneous of degree n}.
Similarly, we have
F and Fn , where 0 is
replaced by~. the antiholomorphic functions. If f, g E
Fn ,
inner product:
126
then we can use the homogeneity to partially integrate the
where r is an open 2disc in the sphere = 1 transverse to the fibers of the Hopf fibration ¢ 2  {0} ~ P1 and mapping isomorphically onto p1 
{a point}. As it stands, the integrand above is not closed.
We can,
however, make it closed, and still have it agree with the above on n2
by multiplying the integrand by n2    b(g) =   g(z 1 dz 2z 2dzj satisfies (a)
ab(g)
In fact, if gEt then n
0

(j
a
(b)
z + z 2 =' b(g) 1 az2 azl
(c)
a a  ' ab (g) zl azl + z 2 az2
(d) for any f E Fn (g,f) = N(n) (e)
•
= 1,
0
(n2)b(g)
Jr
f(z 1dz 2z 2dz 1 )
A
b(g)
b(g) is a exact if and only if g  0.
In particular
F
n
b +
%
H1 (P 1 , 0(n2)).
We note that we can include F and F into L 2 (¢ 2 ,d~) by 1 I(f) = f exp  2.
Then we have the holomorphic and antiholomorphic
Projectors: P+:L
2
~
l(F)
P+f(z) = k
J exp[
~z,z> +  ~z',z'>]f(z')d~' 127
1 1 exp[ ~z,z> +  2 ]f(z')d~'
P_f(z)
Similarly, for any complex vector space V equipped with an hermitian inner product <, >V' we have associated holomorphic and antiholomorphic projectors,
We can think of twistor space ¢ 11
=(:~)and w =(:~)with A((TI,w), (TI',w'))
4
= ¢
2
e
¢
2
with coordinates
an indefinite hermitian scalar product: i(w.
n'
Tl.
w')
and a symplectic form n(p,p') on the underlying JR 8 : n(p,p')
Im(i(w(p) • TI(p') TI(p) • ;;;(p')).
We can, however, put another complex structure on the underlying put 1,; = 
1
12
(11
+ iW} and v = 
1
(11
12
+ iw)
~8
Namely,
as new complex coordinates.
We
then have the positive definite hermitian (w.r.t. the new coord.) inner product B((l,;,v), (l,;',v')) = 1,; • ~· + v •
v'
with the property that
Im(i,;(p) • ~(p') + v(p) • v(p')) = n(p,p'),the same symplectic form.
Let's
call the complex structure with respect to the TI,w coordinates J 0 , and with respect to the
~.v
coordinates, J,
Then we can form, given the positive def
inite inner product, the hilbert space F(R8 ,J,B) of holomorphic functions. There is a (projective) unitary representation of the group of all real linear transformations of R 8 which leave n invariant in this space F which, when restricted to those which also preserve J, is simply the action of coordinate transformations on elements of the BargmannSegalFock
representat~on
F.
[1].
This representation is called When this representation is
restricted to the (covering of) the conformal group, the space
128
F breaks up
into a direct sum of all spin, masszero representations, each occurring but once [2]. Comparing the complex structures J and J 0 we find that on W+ =
=
J = J 0 , while on W_
{v=O} we have J = J 0 •
{~=0},
Therefore, an element of F
when restricted to W+ will be holomorphic w.r.t. J 0 and when restricted to
w_
will be antiholomorphic w.r.t. J 0 •
If we include Fin L2 , we can think
of an element as giving us a holomorphic (resp. antiholomorphic) function on every complex (w.r.t. J 0 ), 2dimensional subspace, V, of twistor space on which A is positive (resp. negative) definite, by using the holomorphic (resp. antiholomorphic) projector associated with the hermitian
(resp.
negative of the hermitian) form on V.
We note that U(2,2), preserving both J 0 and A, has a natural action on the collection of functions g • 6V(f)
= (6 g
{~(f)}.
1 _ 1 (f)) o g
v
Namely, for g E U(2,2), define
While this action does not commute with the
BSF action of g on F, it commutes up to multiplier: Theorem 1 The reason for this multiplier is that we should be dealing with forms rather than functions. Theorem 2
8
Let f be any element of F(lR , B,J).
Define the infinite
collection of functions {~B'C' ••• L'' any number of indices, each index
taking the values 1 and 2}
129
where
~B'C' ••• L'(X)
=
(K(X)~V(X)(f),
w8 ,wc•···;L,j V(X)) for X such that
t i(XX*) is negative definite and K(X) = det ~ (i(XX*))det(I+iX).
Then
~B'C' ••• L' satisfy
a
AA. '~B'C'
(a)
ax (b)
a
axAA'
•••
L' 
a
_AB' ~A'C'
ax
•· •
L'
0
o.
~B'C' ••• L'
Also define ~BC ••• L(X) the
~
satisfy
a
(c)
~
(d)
axAA'
axAA'
a
Theorem 3
a
BC ••• L
~
axBA'
AC ••• L
0
~BC ••• L = 0.
All of the
~A'B'
••• L' and
~AB ••• L
are identically zero if, and
only if, f is identically zero. When f is a monomial in C and v, it is quite straight forward to find the cohomology class which represents the same field information. We will use r P1 Pz rl r2 multiindex notation: ~Pv = cp c 2 v1 v 2 , lrl = r 1 + r 2 , p! = p 1 !p 2 1,etc. Proposition
Iff= acpvr with IPI  lrl = n > 0, and a E t, then ~A ... L 0
for all choices of indices;
~A'
••• L'
=0
if the number of indices is not
equal to n, and ~A' ••• L' (X)
n
t
=
I
nA' ••• nL,(n 1 ,dn 2 ,  n 2 ,dn 1 ,)
Here V(X) is the 2plane associated to the point x E transform.
130
A
a(f)
rcv(x)
~1M
by the Penrose
with respect to the J 0 complex structure and satisfies, again w.r.t. the J 0 complex structure, Cla(f) = 0
(a) (b)
y _j
a(f)
(c)
y _j
Cia
(y=
0 (n2)a(f)
a
()
()
()
TTl ClTTl + TT2 an 2 + wl aw 1 + w2 ()). w2
Therefore a(f), defined on all ofT, represents a cohomology class [a(f)] E H1 (PT ,0 J (n2)).
Moreover [a(f)] is the unique class which
0
represents the same field information as f, by virtue of a
A
Proposition
a
closed (0,1)form on T
satisfying (b) and (c) above which
is ()exact on every complex 2dimensional subspace in T, is actually exact. Reversing the roles of
and v we have
If f = a~pvr with lrl 
Proposition ~A'
~
IPI = n
>
0, then
... L'  0 for all choices of indices.
$A ••• L
=0
if the number of indices is not equal ton.
f
$A ... L =
n:A ... liLs
A
<7i 1 d7i 2  7i 2 dn 1 >.
rcv
= N( Ir I )a~ p v r (v
• v) I r l2 (v 1dv 2  v 2dv 1 ) is a (1,0) form on
T+ satisfying (w.r.t. J 0 ) (a)
as(f) = o
(b)
y
(c)
y _j
_j
S(f)
=0
aS(f) = (n2)S(f)
131
and thus represents an element of H1 (PT+7f ,vJ (n2)). 0
If we want to think of the unprimed information as living on PT, we can resort to defining the form
which lives on PT and satisfies (a)
aa (f)
o
(b)
'Y...J a
(c)
y...Jaa(f) = (n2)S(f).
o
We can define fields
J
$A ••• L (X)
(awA ••• awLB(f))
A
(w 1 dw 2  w2dw 1 )
rCV(X) which again satisfy the field equations, but are not, obviously at least, the ~
same as ljiA ••• L.
Here, awLS(f)
a = a~
~
...J aa(f), etc.
Dealing with the primed spin fields first, we can define a pairing on  0J (n2)). H1 (PT, 0
If a,n' are (0,1)forms representing classes [a],
[ Q']'
let
I
1 <[a], [a' ]> 0 L  r>O N{lrl+n)r!
r r r r II (av 1 av 2a}(av 1 av 2a•) (r;·~)n+ r + 2 1 2 1 2
I'CV(ii)
a ...J aa, etc. This is, in fact, independent of the where again (av )a= 1 avl choice of representatives, a, a'. Moreover, if f = ar;Pvr, f' = a'l;svt with IPI  lrl
=n =
(f,f')F
132
lsi  ltl > O, then
= <[a(f)],
[a(f')l>o·
Although I have no proof of it, it is reasonable to believe that for every f E F(m.8 ,B,J), there exists a cohomology class field information as the
~A'
0:n (f) representing the same
••• L' corresponding to f. n
Given this, then the closure, Hn' under
<,
>o
of the subspace of classes
generated by the a(f), fa monomial of the above type, will actually be 1

contained in H (PT ,0(n2)) since the closure, Gn' under ( , )F, of the space generated by the f's is actually contained in
F.
Then a :G n n
+
H n
coincides on finite sums of monomials with [a(f)] and intertwines the action of U(2,2) on
F and
H1 (PT,0(n2)).
The space Hn is then the space of
normalizable fields. As a final remark, it may be noted that all of the above may be quite easily generalized to other dimensions. twistor space
~
new coordinates k1
H
k
~ ~ ~.~.
~
(PT ,0(nk)).
k
In particular, if we begin with a
, we have U(k,k) acting, the analogous hermitian form A, etc.
The relevant cohomology in this case is
These cohomology classes represent nfold symmetric
fields with k values for each index.
The fields will satisfy formally
identical field equations on the domain {XIi(XX*) neg. definite k x k matrix}. These, and especially the corresponding inner product on cohomology may prove useful in the multiparticle context. For further details on all the above, and proofs for the stated propositions and theorems, please see [2], [3], [4] and the references therein. References 1
V. Bargmann,
Group representations on Hilbert spaces of analytic functions, in Analytic Methods in Mathematical Physics, Ed. R. P. Gilbert and R. G. Newton, Gordon and Breach (1968).
133
2
H. P. Jakobsen and M. Vergne,
Wave and Dirac Operators, and Representations of the Conformal Group, J. Functional Analysis, 24, 1, (1977).
3
c.
The Metaplectic group and holomorphic projectors, to appear.
4
N. M. J. Woodhouse,
M. Patton,
Twister cohomology without sheaves, Twister Newsletter #2.
Acknowledgment This work was supported in part by NSF grant MCS7718723 and an American Mathematical Society Research Fellowship.
C M PATTON Department of Mathematics University of Utah Salt Lake City, Utah 84113
D Burns ·Some background and examples in deformation theory
This report is based on lectures given at the Lawrence Conference.
Their
purpose was to provide some elementary background in deformation theory of complex manifolds and related structures which might be of use in conjunction with the other lectures.
I have tried to sketch the (standard) theory behind
Penrose's nonlinear graviton construction, to give some direction to a general survey.
Formally identical techniques are used in YangMills theory,
of course, and for a more detailed account in that vein, one should check Rawnsley's lectures [8], or the paper [1] of AtiyahHitchinSinger. There are several good references now for the basics of complex manifolds, and their deformation theory, e.g., [4], [10].
The reader may also find
references there to the original papers of Kodaira and Spencer on the subject. Hopefully, these lectures will offer a quick illustration of the basic techniques of the theory, plausible indications of why the theorems should be true, and some examples and methods of computation. The author would like to take this opportunity to thank D. Lerner and P. Sommers for their hospitality during the conference, and editorial patience after the conference.
Thanks are also due them and many of the
participants, who shared their suggestions, opinions and speculations during the conference. §1.
DEFORMATION OF COMPACT COMPLEX MANIFOLDS
To fix ideas, we'll recall that a complex manifold M is a differentiable manifold of real dimension 2n with local complex coordinates z a ... (za, 1 n defined on open sets Ua' ••• ,za)
135
related by complex analytic transition functions faB(z 6 ) = za, faB defined on z 6 (ua n u 6 ) C ~n.
The faB satisfy the compatibility condition
A deformation of the complex
(z )) on U n u 8 n U . i.e., f ay (z y ) = f a..,0 (f ..,y 0 y a y
structure on M is given by making za, faB above depend on additional parameters t =
~
1
k , ... ,t ), so we should have
z (t) a
(1.1)
f
(1. 2)
and
ay
(t,z (t)) = f 0 (t,f 0 (t,z (t))), y a.., ..,y y
and where za(O) = za is a coordinate for our original complex structure on M. Two complex manifolds M, M' are equivalent (M morphism f:M
~
~
M') if there is a diffeo
M' which is complex analytic.
In practice, one rarely deforms a complex manifold in this fashion. Example 1:
Let L(T) be the lattice in t of complex numbers of the form
m + n<, where m, n are integers, and Im T z + m + n< gives a torus M(T) = t/L(T). if and only if <' = Example 2:
>
It
0.
Identifying z with any point
is easy to show that M(T)
~
M(T 1 )
~~ : :. with a,b,c,d integers and ad  be = 1.
Let P(x) be a
homogeneous polynomial of degree d in variables
x = (x 0 , ... ,xm) with complex coefficients, and let VP = {zeroes of P in CFn}. For the generic P, VP is a complex manifold, and if we vary the coefficients of P, we get (generally) inequivalent complex manifolds. Example 3: At 1:_ At I .
136
For 0 < t < 1, l~t At= {z E tit < lzl < 1/t}.
Fort~ t
1 ,
We will see more examples of these kinds below, in §'s 1, 2, 3, respectively. As a first step toward measuring a deformation (1.1), we linearize the
compatibility condition (1.2) by differentiating with respect to t at t
0:
=
i
afaB a fBy
(1. 3)
:1 at azB Here we set Z
cxy
 a fa e Cit
It=O
and interpret it as a holomorphic vector field
on Ucx n UY, and similarly for the other terms in (1.3).
Rewrite (1.3) as
zcxy
(1. 3)
This is the linearized cocycle condition. holomorphic vector fields on
·~i.e.,
If
e
I
denotes the sheaf of germs of
the infinitesimal complex automorphisrns
of M, then the vector fields ZcxB defined on Ua n u8 determine a Cech leocycle on M with values in
e.
The deformation (1.1) will be trivial if, for every t, (1.4)
for then the maps* h (t,z (t)) of the U to themselves patch together to give a ex a an equivalence for each t of Mt with with coordinates z (t). a
Mb•
where Mt is the complex manifold
As before, differentiate (1.4) at t
Clh
at
a
or,
*h a
= 0:
(1.4)'
(1.5)
is holomorphic, for fixed t, with respect to z a (t).
137
aha
where Za' ZB are the vector fields
ahB
a; , a;
defined on Ua' respectively UB.
(1.5) is the coboundary condition. Recall that the first cohomology group H1 (M,0) of M with coefficients 0
"'
is given by the quotient of the space of all 1cocycles, i.e., ZaB on Ua nUB such that (1.3)' holds, modulo the coboundaries as in (1.5).
(For
this definition to be precisely correct, it suffices to assume that the Ua are balls in local coordinatesthe Ua must·be holomorphically convex.) If M is compact, it is a basic fact that Hi(M,0) is a finite dimensional vector space.
1
The computations above interpret H (M,0) as first order
infinitesimal deformations of M modulo trivial deformations.
is the derivative of the deformation at t = 0.
The assignment
We expect there to be at
most dim H1 (M,0) parameters of inequivalent complex structures on M, close to our original one and,in favorable circumstances, this is so.
Before
discussing why this is so, (and in what sense it is so!), let's analyze another example of deformation, useful in YangMills theory. Let E be a holomorphic vector bundle over M. holomorphic m x m matrix valued functions gaB
Thus we're given invertible, defined on Ua n UB which
satisfy (pointwise multiplication) on Ua n UB n UY.
(1. 7)
To deform E, we allow gaB to depend on a parameter t again,
Taking tderivatives of (1.7) at t = 0 gives a a gaB 1cocycle condition on GaB = The derivative of the deformation is
with g
0
ap
(0) =given g
0 •
ap
ae
1
the class of {GaB} in H (M, End(E)), where End(E) is the sheaf of (germs of) infinitesimal automorphisms of E, i.e., linear mappings from E to itself. 138
We'll compute examples of these cohomology spaces below.
Two other
cohomology spaces play an important role in the theory: (a)
0
H (M,0) =space of 0cocycles, i.e., collections of vector fields Z
a
defined on Ua, so that Za ZS = 0 on Ua nUS.
Thus, H0 (M,0) is the space
of globally welldefined holomorphic vector fields, the Lie algebra of the group of global complex automorphisms of M. (b)
H2 (M,0) = vector fields Z defined on U n U n U with aSy a S y
on ua nus n uy
uo' modulo coboundaries, i.e., modulo zaSy such that
(l
on Ua nUS n UY, for some Zas's on Ua nUS.
This space plays the role of
obstruction space to the construction of deformations:
cf. below.
The
"favorable circumstances" mentioned above are when H2 (M,0) = 0, as stated in the following existence theorem of KodairaSpencerNirenberg. Theorem 1 (b)
(a)
If H1 (M,0) = 0, any small deformation of M is trivial.
If H1 (M,0)
+ 0,
H2 (M,0) = 0, then there exists a complex manifold M,
parametrizing a family of complex structures on M, of complex 1
dimension= dimCH (M,0), and such that the derivative mapping described above is an isomorphism of the tangent space of Mat 0 to H1 (M,0). Remarks:
One knows more; namely, in case (b), any small deformation of M
can be induced from the one in the theorem by relabelling the parameters. Much more general results are known about the existence of such "local moduli spaces" M.
Theorem 1 is the simplest of its kind, but, to this date, 2
at least, simplifying assumptions analogous to H (M,0) = 0 have thus far been valid in the cases of physical interest.
139
In case (a) of Theorem 1, M is called rigid. To sketch a proof of Theorem 1, let us look at another way to describe the complex structure on a manifold M, and its cohomology.
This method examines
the CauchyRiemann equations on M directly. If zj = ~ + iyj are local holomorphic coordinates on M, set, as usual, dzj
dxj + idyj
dzj
d~  idyj
and
a£
is the "antiholomorphic part" of the total differential df, and
if and only i f f satisfies the
Cauch~Riemann
a£
= 0
ilf
equations.= 0, j = l, ••• ,n.
a'ZJ
0 k Let A ' denote the space of kth degree exterior differentials a which can be written
=
I:f
and define
aa
a
il'''ik E
_il dz
AO,k+l by
_j
dz One has
a(aa)
solving ila =
dz
i.
.L
0, and locally this is the exact compatibility condition for
B.
More generally, if E is a holomorphic vector bundle on M,
let E ® AO,k denote the space of (O,k) forms with values in E, so a E E ® AO,k i f a looks locally like
140
where {e1 , ••• ,em} is a local holomorphic basis, or frame, for E.
Define
_ik
A
0
o
0
A
dz
Define the kth cohomology group of O(E), the sheaf of germs of holomorphic sections of E, as Hk(M,0(E)) = {globally defined 13 E E ®A O,kl as = 0} modulo{S = irala (globally defined) E E ® AO,kl}. ol
This agrees with our previous definition, using the Cech method. this is because if as
= 0,
one can solve aa
=a
Basically
locally, on neighborhoods
but these local solutions don't patch correctly on overlaps
Ua n
u(l,
u13 .
Carrying out a procedure of successive corrections of solutions on multiple overlaps gives the desired isomorphism. 
0 0
The aoperator from A '
on M, since solutions of af
functions to A0 • 1 "gives" the complex structure 0 determine the complex coordinates on M.
can deform M by perturbing the aoperator.
We
0 1 To do this, let w E T ®A '
where T is the (holomorphic) bundle of tangent vectors involving only the so locally
If w is small, define new CauchyRiemann equations on M by 0,
j
1, ...
,n.
(1. 8)
The compatibility condition for finding n independent solutions of these equations is
aw + [w,w]
0
(1.9)
1111
where locally [w,w]
The NewlanderNirenberg theorem says that this integrability assumption implies the existence of complex coordinate functions which satisfy the new CauchyRiemann equations (1.8). Now think of deforming M by introducing parameters in w, and write formally
Substituting in (1.9) gives successive linear equations awl
0 (1.10)
'liwz = [wl,wl] aw n
p
n (wl' ••• ,wn 1)
for wn in terms of w1 , ... ,wn_ 1 . each p (w , ••• ,w n
1
n1
One checks that 1iPn(w 1 , ... ,tunl) = 0, hence
) successively defines a class in H2 (M,e).
2 role of H (M,e) as "obstruction" space:
We see here the
th we can solve the n  equation
for w if and only if the class of P (w 1 , .•• ,w 1) is 0 in H2 (M,e). n n nConversely, if H2 (M,e) = 0, pick any w1 with aw 1 = 0, and then solve successively for w 's to get a formal !parameter deformation of M.
Conver
n
gence is more delicate, of course, and requires compactness of M. To sketch why a convergent solution exists, use a metric on M to construct 0 1
Hilbert space inner products on T ~A ' complement of the range of
a from
H2
V
be the orthogonal Set V(w) = ~w +[w,w), 
defines a function from R(a)
null space of a:T ® A0 • 2 .... T®A0,3.
range of a:T® AO,l .... T ~ A0 • 2 •
J.
T® AO,O toT® A0 • 1 •
w E T ® A0 • 1 , and note aV(w) = 0, so Ker('li)

and let R(a)
I f H2 (M,e)
The derivative of
V
at w
o,
.J.
to
Ker(a) is the
0 is just
V'(O) =a, which maps
R(a)~

the set of w near 0 in R(3)
onto Ker(a). ~
By the implicit function theorem,
with V(w) = 0 is a manifold with tangent space
at w = 0 given by the null space of V' (O) = a.
H1 (M,0).
This is identified with
The compactness of M and the ellipticity of the asystem are used
to justify the analysis above. Remarks:
Some modifications are required to drop the assumption
(A)
H2 (M,0) = 0 in the above proof.
Then the solutions of V(w) = 0 as above
don't necessarily form a manifold.
This extension is due to Kuranishi.
Grauert has extended the theorem to the case where M is allowed to be a singular space. (B)
With minor modifications, the method described here works for
deforming complex vector bundles over a fixed M. Finally, lP.t us compute some easy examples. (1)
The Riemann sphere = O:P 1 •
~
is the complex line bundle over G:P 1
whose holomorphic sections over an open set U are those holomorphic functions fin the homogeneous coordinates x,y for U satisfying f(AX 1 Ay) = Anf(x,y). Thus,
Ho(ctPl
'
0(~))


{0,
if n < o space of homogeneous polynomials in x,y of degree n, n ~ 0.
To calculate H1 (G:P~ O(HGn)), we use Serre duality*to conclude
*' ·oerre duality says the spaces Hi (M, 0 (E)) and Hni (M, 0 (E~An ' 0 )) are dual; where M is compact, E is a holomorphic vector bundle, E* is the dual bundle and An,O is the bundle of complex "volume forms" adz 1 ~ ••• ~dzn. In terms of j j 0 · 1 n s1  sni f ~dz ~dz ~ •.• ~dz orms, if e ® dz 1 ~· •• ~dz i is in E ® A • \ and e*®dz M
are annihilated by
a,
Paired to the number
••
the classes in Hi(M,0(E)) and Hni(0(E* ® An,O)) are
f (e,e*)dz
1
~ ••• ~dz
2
1
~dz
n
~ ••• ~dz
•
M
11,3
The sheaf e is just O(H
02
1
1
) • so H (CF • e) = 0 and O:F
1
is rigid.
(O:F
1
doesn't even have discrete changes of complex structure, by the uniformizaWe should also note that H1 (0:F 1 , 0)
tion theorem.)
the sheaf of analytic functions on O:F
1
=
0, where 0
=
0(H00 ) is
For any holomorphic line bundle L
on any complex manifold, the infinitesimal deformations of L are given by H1 (M,0), since 0
=
O(End(L))any linear mapping of a line bundle to itself
is just multiplication by a holomorphic function. line bundles on O:F
1
are rigid.
Since H1 (0:F 1 , 0) = 0, ~
(These are actually only the H
given
above.) (2) e
Tori T
=
0:/L(t).
ForT, the vector field
and dimO:H 1 (T,e)
= 0,0(A1 • 0 ) = 0
=
dimO:H 0 (T,0)
a:
= 1,
is nonvanishing, so by duality.
We have
already seen that one parameter, namely t in example 1 earlier in this §. M = O::n>1
(3)
X
Here e = el ~ e2. where ei is the sheaf of vectors
a::n> 1 .
th factor, i tangent to the i1
1
= 1,
The Klinneth formula * tells us
2.
1
11
01
H (M,e) = H (M,e1 ) + H (M,e2 ) = two copies of H (CI:F , e)® H (CI:F ,
+ H0 (o::n>1 , e) ® H1 (o::n>1 ,
)
0) = 0, by the earlier example.
*The Klinneth theorem says that, if Ei is a holomorphic vector bundle on i
= 1,
2, and E is the bundle
PfE1 x
p~E 2
on M
Mi,
factor),
=HI
x
Hz
(where pi:
M ~ Mi is the projection onto the
k
i
R.
Then H (M,O(E)) = k!t H (M!,0(E1 )) €) H (M2 ,0(E2 )).
The map is
given, in terms of differential forms, (al ® wl) x (a2
0 w2)
~ (p{al
®
p~a2)
® Pfwl
~ p~w2 •
Here aj is (locally) a differentiable section of Ei, and wj are (O,k) or (0, Jl.) forms.
.th
1.
§2.
RELATIVE DEFORMATION THEORY
Suppose now that M is a compact, complex manifold, and that M C M, where M is another complex manifold, not necessarily compact.
We have a short
exact sequence of holomorphic vector bundles on M: 0 +
where TM
TM + T1M M + N + 0
=
(2.1)
bundle of holomorphic tangents to M, TMIM
at points m E M, and N = normal vectors to M in M.
=
tangents to M, tangent
We get a long exact
sequence from (2.1): ~
...
i
Here we've omitted mention of MinH (M,0M), etc., and used 0M for O(TM), ~~M for O(TMIM).
At the infinitesimal level, deformations of M inside M
will be measured by H0 (0(N)), a normal holomorphic vector field along N giving a first order "motion" of M in M.
Some fields come via a from bolo
morphic fields on M, so they give trivial deformations of the "abstract" 1
manifold M when considered in H (0M) via 61 •
The obstructions to realizing
a given cr E H0 (0 (N)) as the deirvati ve of a deformation in M lie in H1 (0 (N)) , and 62 of these obstructions measures whether one may deform the structure on M by moving M "outside" of M. Let us first sketch the basic setup, in a Cech formalism, as follows: We'll be given a covering {Ua} of M, and we'll want varying coordinate functions za(t), and transition functions faa(t,za(t)) satisfying (1.3), and imbeddings ha (t,za (t)) of Ua into M, holomorphic at "time" t with respect to
145
za(t), with ha(O,za(O)) the identity imbedding.
These are all to satisfy the
condition
(2.2) aha As before, ~should be the derivative of the deformation, but
read only modulo tangent vectors to M, since moving M tangent to itself doesn't change the structure of M as complex submanifold of
M.
At t
= 0,
we get:
(2.2)'
=
ah
aea
mod tangents to M, aha
since ha(O,za(O)) is the identity.
ae•
(2.2)' implies
gives a globally defined normal vector toM in
M,
read modulo tangents,
i.e., a section
oE H0 (0(N)).
Conversely, given such a o, it determines ha's, to first aha 'V 'V order, in the normal direction. Once we realize (locally) ~ = Z , Z a af at a a aa holomorphic section of ~ M' we've determined~ at t = 0 in (2.2). (This
I
is just computing
o1
of o.)
The first obstruction occurs when we try to determine the second order a2 expansion of ha in t. Taking 2 at t = 0 of (2.2) gives at a 2 ha
a 2h a a 2h_a+ _ 2 R.at 2  at 2 aza at
ar9. aa
at 
(2.2)"
0,
HR. a 2h a ae Here the expression "(2) = . zaa Jl. at azaat one cocycle for O(N), i.e., when read modulo tangents to M.
mod tangents to M.
determined by our first order choices already made.
on Ua n ua is a Note that it is
If H1 (0(N))
= 0,
we can
~~~)
write
a2 h a at 2
~( 2 ) + ~( 2 ) mod tangents, and satisfy (2.2)" by choosing a
13
"(2) =
za .
The analogue of Theorem 1 is: Theorem 2 (Kodaira)
If H1 (0(N))
= 0,
then there exists a d(complex)para
meter family of deformations of M inside
M,
where d
=
dimCH 0 (0(N)).
The
derivative mapping is an isomorphism of the tangent space to the parameters at 0 with H0 (0(N)). Remarks:
The nonlinear operator proof of this statement is harder to set UJ
than the previous one
V
the
(Theorem 1).
This is because the operator, unlike
of §1, must involve to all orders the imbedding of Min M.
v
already seen this in the first Cech obstruction:
~~~)
We've
the second derivatives in
depend on the imbedding, although H0 (0(N)) and H1 (0(N)) depend only on
the (first order) normal bundle N over M. The nonlinear operator method does work, and has the advantage of applying to other situations, for example the parametrization of nearby minimal surfaces in the calr.ulus of variations, where no cohomological or sheaftheoretical arguments are available. Douady has given a much more general form of this theorem. Of importance to us is the following corollary, which deals with a "doublyrelative" situation:
suppose we now continuously deform
some (small) complex parameter s. M, M in s
Ms ,
H (M,O(N))
=
Do we have a continuous deformation M of s
and if M has a dparameter family of deformations in
have such a family in 1
Mto Ms for
0.
M ? s
M,
does M s
The answer to both questions is yes, if
This follows as an immediate corollary of Theorem 2 if To see this let D be a
deformation of M. 0
+
N
+
M=
u M be the "total space" of the s sED We have a short exact sequence of normal bundles on M:
small parameter ball in tr. and let
NM. M+ NMl M. M
+
(2. 3)
O.
where NM.M is the normal bundle of M in M. and NM.MI H is the normal bundle NM.M jM is isomorphic to the tangent bundle
to M inside M. restricted to M.
TD of D pulled back to M. so is trivial. and O(NM.M~~~ is isomorphic to $ 0. r copies
In cohomology. then. we get
Since the outer terms are assumed theorem applies to Min M. r =dime H0 (0). r
o.
1
so isH (O(NM.M)).
Hence. Kodaira's
But dim H0 (0(NM.M)) = dim H0 (0(N)) + r.
Since we already know d =dim H0 (0(N)) of the parameters
give the deformations in
M.
we know that the remaining r must give deforma
tions of M into all of the nearby
Ms • Further. each Ms • for s small. must
contain a dparameter family of deformations of M.
(Note that each small
deformation Mt of M must lie entirely in one of the Ms• since Mt is still compact and the usual coordinate functions z 1 ••••• zr on D must be constant on Mt• when pulled back to M.) As examples of the above. let us make some of the basic calculations used
in the construction of nonlinear gravitons or the "nonself dual twistor transform." (1)
For the nonlinear graviton in curved twistor space. we want to let
 = an open neighborhood of that line. M = a line in IC1P 3 • M
The normal
bundle N will be twodimensional. and since M is the intersection of two planes E1 and E2 in G:1P 3 • N is the sum of two line bundles Ni of normal vectors to M in Ei. i = 1. 2. 148
Each Ni is just the H = H01 of §1. since
Ei
=
C1P 2 ,
and in homogeneous coordinates [x,y,z], if M is given by x
a
zay is tangent vector on where z
=
x
= 0,
~1P
2
= 0,
which is not tangential to M along M except Since only linear functions on t 2
i.e., at one point in M.
1
have one simple zero when thought of as sections on M = C1P , Ni must be H. Hence,
The set of sections a E H0 (M,0(N)) which vanish somewhere on M form the complex light cone in ¢ 4 • If we deform the ambient M, we first note that H1 (M,0) Since H1 (M,0)
corollary to Theorem 2 applies.
M inside M = U M must be isomorphic to M. s
Nt
= normal
s
so that the
all the deformations of
Hence, the normal bundles
bundle of Mt in Ms (s determined by t) give a deformation of the
original N on the fixed manifold M. N:
= 0,
= 0,
End(N) iG the bundle (H
Let us compute the deformations of that
$H)~ (H([email protected] H®l), so
Hi (M,O)
={
4 copies
¢8'
0, i
i
= 0
=1
Hence, N is also rigid, and the infinitesimal incidence relation {a E H0 (Mt ,O(Nt>>l o vanishes somewhere on Mt} is once again a nondegenerate quadratic cone, which varies holomorphically with t.
This gives the bolo
morphic conformal structure in Penrose's construction. For the nonselfdual transform, we consider C1P 3
(2)
&eneous coordinates xi, yi, i 
M C 11:1P
3
x
11:1P
3
given by I:xiy i
= 1, ••• ,4, =
0.
x C1P 3
with homo
on the respective factors, and
Let ~1P
1
x
¢1P
1
= Mc
Q,
e.g., {x1
= x 2 = 0}
x
{y 3
= y 4 = 0}.
The normal bundle N is harder to
compute exactly here, and we'll just compute its cohomology using exact sequences.
First, let's denote by O(k,t) the sheaf of sections of the line
bundle on M obtained by tensoring ~ on the first factor with ~ on the second.
We've an exact sequence of normal bundles again: (2.4)
where N2
normal bundle of M in G:lP 3 x ClP 3 ,
N3
normal bundle of M in CP 3 x CP 3 ,
restricted to M.
0(N 2 ) is just 0(1,0) + 0(1,0) + 0(0,1) + 0(0,1), arguing as before.
0(N 3)
is just 0(1,1), basically because the form Exiyi is quadratic, linear in the x's and the y's. 0
+
o
H (0(N))
+
Consider the cohomology of (2.4):
o
H (0(N 2 ))
ao H (0(N
+
3 ))
cSl
+
H (0(N))
+
1
H (0(N 2 ))
+
•••
(2.5)
..
In the following calculations we use the Kunneth formula
E& 4copies
H0
o(H))
x
H0
= G:8
H1 (0{N 2 )) = 0
H0 (0(1,1)) = H0 (tP1 , O(H)) ® H0 (tP1 , O(H))
1
To apply Kodaira's theorem, we want H (O(N))
cc 4
= 0,
and to recover complex
conformal Lorentz manifolds we again want H0 (0(N)) = a: 4 with a holomorphically varying cone of infinitesimal incidence cones which are nondegenerate quadratic cones. in (2.5).
Both dimension counts work out if we check B is surjective
If [x,y] and [u,v] are homogeneGus coordinates on the first and 0
second faetors of M, then an element a of H (0(N2 )) is identified with a 150
4tuple of linear forms (f1 ,f2 ,g1 ,g2 ), the fi in x andy, the gi in u and v. An element T E H0 (0(N 3)) is a quadratic form in x,y,u,v, linear in x,y and
in u,v.
The map B is, up to linear change of the homogeneous variables,
given by
This is As
obviously surjective.
noted earlier, M is rigid, and so analogous to Example 1 just above, we
now have only to show the normal bundle N is rigid on M.
We won't carry that
out here, but it can be done by comparison with N2 and N3 again. Remarks:
In the full nonlinear graviton construction one wants to know not
just the conformal structure but rather the actual set of deformed curves or surfaces, and their intersection relations.
For this we refer the reader
for examples to either the paper of CurtissLernerMiller [2] or Ward [9]. For more on Example 2, cf.the papers of IsenbergYasskinGreen [3] or Witten [11]. §3.
CONCLUDING
REMA~S
The deformation theory outlined above is sufficient for setting up the local formulation of nonlinear gravitons.
The global theory, which appears not
to exist now, is much more difficult, especially since the problem is not very precisely formulated, so far as I understand the state of the art. Roughly speaking, one wants global deformations of "all" of lP 'D.'+ or lP 'D.'to give a truly global nonlinear graviton.
I say "all" mainly because it +
is extremely unclear what one wants to do with the boundary N of lP'D.'. In general, very little in the way of existence theorems is known for deformations of open complex manifolds M with boundaries aM.
Various works 151
of Hamilton, Kiremidjian and Stanton appear to have the fault that the boundary conditions imposed appear either too strong for what is needed in twistor theory, or else the assumptions on eigenvalues of the Leviform of the boundary are not met by N. One might try sticking to the compact manifold N and deforming it as a CRmanifold, i.e., deforming its partial complex structure it inherits from the ambient lP 'Il' •
One loses ellipticity, however, in the resulting system
the ab systemof equations on N.
Further, one does not even know in the
most favorable cases (positive definite Leviform, and dimension
~
5) whether
the integrability theorem analogous to the NewlanderNirenberg theorem is true.
Further, still, theN of twistor theory isn't even a favorable case!
It seems to me (which is not much of a recommendation for what follows) that the global formulation of the problem required here is very akin to the perhaps simpler problem of superposing "right" and "left" solutions of the linearized field equations, and the "norming" of such fields to give some sort of probability amplitudes.
The boundary conditions needed will hope
fully be determined by what is wanted physically, with the aid, perhaps, of some insight into particular solutionsmuch like the origins of twistor theory in the expression of solutions of field equations by contour integrals. Finally, I apologize for not mentioning deformations of singular varieties enough to help introduce the reader to the "Euclidean gravitons" constructed by Hitchin. References 1
2
152
M. F. Atiyah, N. Hitchin, I. M. Singer, Selfduality in fourdimensional Riemannian geometry, Proc· Roy. Soc. Lond., 1978. W. D. Curtis, D. E. Lerner, F. R. Miller, Complex ppwaves and nonlinear gravitons, to appear in GRG.
3
J. Isenberg, P. Yasskin, P. Green, NonSelfdual gauge fields, Preprint (U. of Maryland).
4
K. Kodaira,
Stability of compact complex submanifolds, Am. J. of Math., 1963.
5
K. Kodaira and J. Morrow,
Complex Manifolds, HoltRinehart, 1972.
6
M. Kuranishi,
New proof of the existence of deformations of complex structure, in Complex Analysis, Springer, 1965.
7
R. Penrose,
Nonlinear gravitons and curved twistor theory, Gen. Rel. Grav., 1976.
8
J. Rawnsley,
Differential geometry of instantons, preprint (Dublin Institute for Advanced Studies).
9
R. S. Ward,
A class of selfdual solutions of Einstein's equations, Proc. Roy. Soc. Lond., 1978.
10
R. 0. Wells,
Differential analysis on complex manifolds, PrenticeHall, 1972.
11
E. Witten,
An interpretation of classical YangMills theory, preprint (Harvard University.) D BURNS Department of Mathematics University of ~fichigan Ann Arbor, ~Uchigan 48104
153
ETNewman Deformed twistor space and Hspace
In this note, we will describe how the space of solutions of a second order differential equation, the socalled good cut equation, defines two interesting manifolds, deformed twistor space [1,2] and Hspace [3,4), and further how these solutions induce on these spaces in a natural manner a Kahler and a complex Riemannian metric, respectively. 2 2 We begin by considering S x S as a complex manifold coordinatized by the complex stereographic coordinates
~
(~.~).
(The sphere metric for an S
2
in these coordinates is
(1)
with bar denoting complex conjugate.)
Actually we will be concerned not with
2 2 the entire S x S but with the antiholomorphic strip defined by an open neighborhood around the line
~
~
=
~.
thickening of the real sphere so that its values close to
~.
i.e. we will be considering the complex ~
~.
though still independent of
has
2
This region will be referred to as SC.
The good cut equation [3] for the holomorphic function
s 2cc )
~.
~
Z(~.~)
(defined on
has the form
a
~ 2..!_ _ o ~ ~ (1+~~) a~Za (z,~.~)
(2)
with a 0 an arbitrary spin 2, conformal weight 1 holomorphic function in the
154
three variables
'U
Z,~.~.
(For the meaning of spin and conformal weight, see
Appendix A. ) (As an aside, we mention that the good cut equation has its origin in a physical question [3).
In Minkowski space a light cone emanating from a
point has a vanishing asymptotic shear or distortion.
The question naturally
arose, does an arbitrary asymptotically flat physical spacetime have any null surfaces which resemble Minkowski space null cones in the sense of having a vanishing asymptotic shear?
The answer in general is nohowever if
the physical space is analytic in the neighborhood of null infinity, by thickening the real space into the complex, one can find complex null surfaces which are asymptotically shear free.
In fact it is the good cut
equation which determines these surfaces, with every regular solution on
2
S~
corresponding to such a surface.
The
'U
a 0 (Z,~.~)
'U
Z(~.~)
in the good cut equation
is a measure of the strength of the gravitational radiation associated with the asymptotically flat spacetime.
For Minkowski space, where there is no
radiation, one has a 0 = 0.) There are two points of view we can adopt towards solving the good cut equation. a)
It can
be thought of 'U
~
IsJ
as an ordinary 2nd order differential equation with
as a parameter, with the general solution thus depending on two 'U
'U
arbitrary constants (say a and B) plus b)
'U
'U
'U
~.
i.e. u
'U 'U
'X
= Z(~;~,a.~),
or
'U
the a and B can be thought of [3,4] as functions of ~ which must be 'U 2 chosen subject to the condition that Z(~.~) be holomorphic on S~.
These points of view can be illustrated with the goodcut equation from Minkowski space, i.e. with a 0 = 0.
a
"' 2acz a
a~ (1+~~)
0
Thus
(3)
155
and hence
(4)
which is the general solution
"'a "'a
from the first point of view.
If one now lets
a~+b (5)
c~+d
then
z
ar;r;+br;+c~+d
(6)
l+r;~
with a,b,c and d being constants. It is not difficult to convince oneself that the choice (S) yields the most general solution to (3) which is holomorphic on
2
5~.
For future reference we point out that (6) may be rewritten in the forms
z
(7)
with the sum going over the first four spherical harmonics and
z
(8)
"' "' 1 (l+r;r;,r;+r;, "' "' r;r; "' , R.a = 2~ (l+r;r;) i , l+r;r;)
(9)
0
1
2
3
.
(x ,x ,x ,x ) are linear comb1nations of the a,b,c,d and R.
a
can be thought of as a complex Minkowski space
null vector for each value of r; and~
As r; and~ move over 5 2 x 5 2 , R.a
spans the entire complex Minkowski space null cone.
156
Note that from point of view a) the general solution to (3) depends on three arbitrary complex parameters
'V'),o'V
and the solution space thus forms a
a,~,C
three complex dimensional set while from b) the solution depends on four a arbitrary complex parameters x and thus forms a four complex dimensional manifold.
We will see that this result holds true for the general good cut
equation with point of view a) leading to deformed twistor space and b) to Hspace. The two points of view may be visualized as a) a line and b) a surface (ruled by the lines) in a complex three dimensional space.
uf
See figure 1.
uf , I
''
. '
c
( 0)
A Point
In
J>7
(b) A Point m
}(space
Figure 1 Twistor Space From the first point of view the general solution of (2) can be
"'
"'
written u = Z(C;C,a,B) "' "' where a and
"'a can
position and slope of the curve at
c = c0
t he curve lies.
be, for example., the "initial"
"'
and C defines the plane in which
"' ':It "' are to be thought of as the local coordinates of the a,~.c
three complex dimensional manifold of these curves.
Knowledge of the
solution u = Z(C) allows one to find the "final" position and slope at C = Which could also be used as local coordinates.
c1
The solution itself thus acts
157
to define the transition functions for the manifold.
This results in this
space being identical to the deformed projective twistor space of Penrose, PT. Deformed twistor space
T (1,2,5] can be defined by introducing homogeneous
coordinates
w2'
"' r, w3'
and the transition functions induced from PT.
Note that u
a'
Z(r,;,w
) is
homogeneous of degree zero in wa' We consider also the conjugate equation to the good cut equation, namely
a a~
"' 2 (l+tt)
a"' = "o 'll "' a <~.t.t)
~z
,
(10)
ar,;
"' "' where "o a (Z,r,;,r,;) is obtained by analytically extending the complex conjugate
of a 0 away from the surface where
"'1,; = 1,;.
The solutions
u = "'"' Z(r,;,za ) thus define
a manifold
(11)
t
conjugate to
have taken a slight liberty
her~
T.
(It should be pointed out that we
with the conventional notation.
What we
have here called twister space T is conventionally called dual twistor space
"'
and T is the conventional twistor space.) We will now define on the product manifold
Tx
T, in a neighborhood of the
diagonal
wa'
za
(the neighborhood being referred to as CK and the diagonal asK), the function a') = irr(l+r,;r,;){Z(r,;,z)"' "' "'"' a I(z a , w Z(r,;,wa' )},
158
or a
a'
l:(z ,w
3' a 2 2' 3 3' ~ w ) = i(z w + z w ) {z( zt•Z ) w
which is obviously homogeneous of degree (1,1) in z
(12) a
a'
and w
1:, by being
considered as the Kahler scalar and hence defining on K the hermitian metric
(13)
with
J>' makes K into a Kahler manifold.
The signature is (++).
The Kahler scalar, which is not usually a geometric quantity, is so in our case.
It measures (with an appropriate scaling) the difference in u values
=
between a curve u common generator
~
~
Z(~.~ 0 )
=
';\;"
~O'
~
~
and its conjugate curve u
=
=
~~~
Z(~.~ 0 )
along the
~
~0 •
The Kahler manifold obtained in this manner has many remarkable features, the most surprising being that the Ricci tensor vanishes and that the Weyl tensor has only three independent nonvanishing componentsthese components being identical to the radiation components of the original physical space Weyl tensor. In the special case of a 0 a
a
l:(z ,z )
=0
one has
13 20 31 = z o2 z + z z + z z + z z
(14)
the standard flat space twistor norm.
159
Hspace Penrose and Atiyah [3] have shown that the solutions to the goodcut equation for sufficiently small a 0 , which are holomorphic on
2
S~,
complex dimensional manifold, which is referred to as Hspace.
form a fourThe solution
can thus be written as "' a ) Z = Z(r;,r;;z
(15)
with the four variables za parametrizing the space of solutions or acting as the local coordinates of the Hspace. With no attempt at proofs we will now summarize some of the remarkable properties of an Hspace. 1)
An Hspace has a complex Riemannian metric naturally induced on it by a
a
gab(z )dz dz
b
(16)
with
dS =
2
i
dl;/\d~
"'
(1+1;1;)
2 ,
(17)
the area element on the complex unit sphere, and (18)
the integration being taken over the real sphere or, to avoid singularities, an appropriate deformation of the real sphere into the complex.
We wish to
emphasize that it is not at all obvious that the integral in (16) defines a quadratic metric.
160
It does however follow from the properties of the good
cu~
equation.
(The metric obtained from (16) with the solution in equation
(6) is the Minkowski metric.) 2)
An Hspace is Ricci flat, i.e.
o. 3)
(19)
The Weyl tensor of an Hspace is selfdual, i.e.
cabed *
(20)
In other words for each real asymptotically flat physical spacetime (and hence for each a 0 ) , we can generate via the goodcut equation a complex selfdual solution of the Einstein equations. If the conjugate goodcut equation is used, one obtains in a similar ~
manner an Hspace with an antiselfdual Weyl tensor. Remembering that the starting point of this investigation was the search for asymptotically shear free null surfaces in an asymptotically flat spacetime, it is natural to inquire into the asymptotic flatness of the Hspace and to see if one can construct the Hspace of an Hspace.
It turns out that
not all Hspaces are asymptotically flata necessary but by no means sufficient condition for asymptotic flatness being
If however we assume that an Hspace is asymptotically flat then one has these further results: 4)
The radiation field of the Hspace (i.e. the asymptotic behavior of
the Hspace Weyl tensor) is identical to that of the original physical space. 5)
The Hspace constructed from an Hspace is the original Hspace.
161
6)
~
~
The Hspace of an Hspace (or the Hspace of an Hspace) is complex
Minkowski space.
As a final point we wish to briefly discuss a version [6] (due largely to G. Sparling [7]) of the AtiyahWard approach to self or antiself dual YangMills theory which closely resembles the Hspace construction. Consider an (n x n)matrixvalued holomorphic function with u
= x a t a (~.~) ~
(see Eq. 8), where the variables x
a
~
A(u,~.~)
2
on St
are to be interpreted
as standard Minkowski space coordinates; and consider the following differential equation for the matrix G(xa.~.~):
(1+~~) ~ a~
GA .
(21)
We claim (with no proof given here) that a regular solution (regular in the sense of holomorphic in
2
S~)
generates a selfdual
field and that all such fields are so generated.
GL(n,~)
YangMills
Specifically the vector
potential is given in terms of G by G Gl + &"hl  hffl 'a a a
(22)
with
h
fi'R.
(23)
a
Though it is not obvious, one can show that A is independent of a and yields a selfdual YangMills field.
162
~
and
~
~
It is Eq. (21) that is the analogue for YangMills theory of the goodcut equation. Appendix A If crosssections of an arbitrary line bundle over homogeneous functions of n A f(An ,
s2
are represented by
A and A' n (A = 0,1) such that
A' ws w+s A A' An ) =A A f(n , n ), then the spinweights and conformal
weight w function n( s,w )(c,~) is defined [8,9] by
(A.l) P
1
=2

(l+CC) •
Using the homogeneity of f one also has f(nA, ;A') = (nl)ws(n l')w+s Pwn(s,w)(C.~) 'II
0

(A.2)
 O' 'II
c= 1 , c = =I" n
n
These bundles classified by their homogeneity degrees or by s and w, though distinct from the point of view of representation theory, are not topologically distinct.
In fact the first Chern class which characterizes
(topologically) any smooth line bundle over S c
2
is given by
= 2s.
(A.3)
!Ppendix B We will here list a series of problems and questions for which we do not know the answers. 1)
To me the most important question is what relationship (if any) does
the theory described here have with physics?
Penrose [1] argues very
convincingly that an Hspace should be viewed as a nonlinear graviton, though how to make a quantum theory of gravity from this idea us.
still eludes
In addition Hspace theory appears to be remarkably well adapted to the
theory of equations of motion [10,11], though again we are at an impasse. 2)
We would like a method of producing solutions to the good cut
equation.
Though G. Sparling [12] has produced a finite set of solutions and
R. Lind [13] has invented a formalism to produce another set, these are quite limited both in number and interesting qualities.
One can on the other
hand produce [14,15] solutions to the selfdual vacuum Einstein equations with considerable ease, though the general solution is not known. 3)
We would like a good useful definition of an asymptotically flat
Hspace. 4)
The definition in current use [10] is rather awkard and unattractive.
What are the conditions on the u 0 in the good cut equation which
guarantee an asymptotically flat Hspace?
"'
5)
Given [6] an
A(u,~.~),
6)
What are the
condition~
how can one construct regular solutions to (21)? on A such that the YangMills field is
"asymptotically flat?" 7)
Are all Ricci flat, selfdual spaces in some local sense equivalent
to Hspaces, i.e. are they derivable from a good cut equation?
If so, bow
does one find the a 0 ? References 1
R. Penrose,
2
E •. T. Newman, J. Porter, K. P. Tod, Gen. Rel. Grav. to appear.
3
R, Hansen, E. T. Newman, R. Penrose, K. P. Tod, to be published, Proc. Roy. Soc.
4
M. Ko, E. T. Newman, K. P. Tod,
164
Gen. Rel. Grav. 7(1976), 31.
in "Asymptotic Structure of SpaceTime" ed. P. Esposito and L. Witten, Plenum Press, (1976).
5
M. Ko, E. T. Newman, R. Penrose,
J. Math. Phys. 18, 58(1977).
6
E. T. Newman,
On Source Free YangMills Theories to appear in Phys. Rev.
7
G. Sparling,
Preprint Univ. of Pgh.
8
w.
J. Math. Phys. 19, 874(1978).
9
A. Held, E. T. Newman, R. Posadas, J. Math. Phys. 11, 3145(1970).
D. Curtis, D. E. Lerner,
10
M. Ko, M. Ludvigsen, E. T. Newman, K. P. Tod, The Theory of Hspace, preprint, University of Pittsburgh.
11
M. Ludvigsen,
G. R.
12
G. Sparling, K. P. Tod,
Preprint, University of Pittsburgh.
13
Private correspondence
14
J. Plebenski,
J. Math. Phys. 16, 2395(1975).
15
A. Janis, W. Fette, E. T. Newman,
J. Math. Phys. 17, 660(1976).
G.~.
357(1977).
Acknowledgment Research supported by the National Science Foundation.
E T NE~T Department of Physics University of Pittsburgh Pittsburgh, Pennsylvania 15260
KPTod
Remarks on asymptotically flat .Y~spaces
1.
INTRODUCTION
In this article, I wish to discuss the notion of asymptotic flatness for a complex solution of Einstein's equations.
I shall be concerned only with
leftflat spacetimes, where the Weyl tensor is selfdual, which arise as Hspaces [14]. To define·a complex null infinity El for a complex spacetime, one might seek a structure analogous to the tl of Minkowski space.
However, nothing
as "big" as this can be expected because a holomorphic metric with curvature will necessarily have singularities which,in particular, will occur at null infinity.
Further, a definition in terms of a boundary will not work since
tl, being six real dimensional, is not a boundary of the original eight real dimensional manifold.
(This is related to the fact that the null cone does
not separate the tangent space; there is no definition of "timelike" or "spacelike" directions in a complex spacetime.)
There is also the
possibility that a number of different nonsingular regions may be added as limit points, resulting in a number of possible asymptotes or parts of El. However, when dealing with an Hspace, there are extra structures which can be employed in a definition of asymptotic flatness to avoid these pitfalls. I shall begin by briefly recalling the salient points of the Hspace construction and giving some motivation for why one cares about asymptotic flatness.
The method of development is to proceed naively, introducing what
seems necessary into the definition.
Then with the aid of an example in §3,
it is possible to polish up the definition.
166
The Hspace construction, [14],
begins with the real future null infinity~!: of an asymptotically flat spacetime M and thickens it slightly into the complex, obtaining ti:. Geometric quantities on
+ notably the asymptotic shear o 0 (u,~.~) which
~IM,
determines the asymptotic gravitational radiation field in M, are assumed to
+ be analytic, so that they extend to holomorphic functions on tiM.
This
effectively restricts the distance into the complex which the thickening of
+ can go.
:RI~1
The Hspace of M [1] is defined to be the four complex dimensio1
+ that is of regular solutions to the "good cut manifold of "good cuts" of tiM, equation" 2
'\,
It Z(~.~)
(1.1)
a
'\,
Here ~ is roughly a~ and ~ is the complexification of ~. a complex coordinatE independent of
~.
'\,
Regularity means that the solutions
on the Riemann sphere of
~
when
'\,
~
=

~.
Z(~.~)
are nonsingula
(See [14] or the article of Newman
in this volume for further details). He write the complete solution of (1.1) as Z(z
a
'\.
.~.~)
where the z
a
are four
parameters on which the solution depends i.e. they are coordinates on the Hspace.
It is possible to define a quadratic metric and curvature tensor
for Hspace in terms of the "good cut function" Z.
The curvature satisfies
the Einstein equations automatically and further is selfdual (or leftflat)
[14].
Thus o 0
,
the data for the original real solution, also determines a
leftflat, complex solution.
A picture of the Hspace construction as a pro
jection of the selfdual part of the radiation field of the original real space~time
suggests itself.
For this to be appropriate, we need to know that
the construction is idempotent i.e. that the Hspace construction applied to ~n
Hspace reproduces that Hspace.
Since the construction is only defined
on asymptotically flat spacetimes, we therefore need to know when the HspacE is asymptotically flat.
167
In §2, following remarks of R. Penrose, I point out that there is a canonical identification between ti: and what one would want to call «IH' the null infinity of Hspace.
Thus one knows what tiH is, the question remaining
is whether it can be smoothly attached to Hspace as a set of limit points. Because of this identification, one also knows what the "real" points of tiH
+ are, namely the ones corresponding toRIM. + of a real RIH
Thus tiH appears as a thickening
The definition of asymptotic flatness for Hspace therefore
begins with the assumption that there is a "real slab" (a term due to M. Ludvigsen) which is the complex thickening of a "real slice" and on which the metric is nonsingular.
The metric itself is not real on the real slice.
The real slice is simply defined by a reality condition on the Hspace coordinates and is a real submanifold which should extend to infinity.
The
real slice can be moved around inside the real slab. For tiH to appear as a thickening of the boundary of the real slice, we need to ensure that null geodesics from the real slab escape to infinity. One would not expect all null geodesics from a point in the real slab to escape to infinity, since some of them will necessarily leave the slab and may encounter singularities.
Instead, the second stage in the definition is
to assume that null geodesics in directions "close to bei.ng real" get away. Specifically, we demand that in the S
2
x S
2
of possible null direction at a
point there is an antiholomorphic "diagonal" to infinity.
s2
of null geodesics escaping
(Antiholomorphic for the following reason: if n is a stereo
graphic coordinate on the first
s2
unprimed spinors AA on the second
labelling primed spinors TIA' and
s 2 then we want a relation
n = f(~)
the role of complex conjugation of spinors in a real spacetime. defines the analogue of real null directions.) is still a little stronger than necessary. 168
~ labels to play
This then
We shall see in §3 that this
In an Hspace there is a preferred
affine parameter on null geodesics, defined up to an additive constant. Consequently one can consider moving real affine parameter distances along a null geodesic, and we need only demand that these real directions avoid singularities and get to infinity. With these conditions on an Hspace (and one other which is probably redundant) it is possible to prove the idempotence of the Hspace construction ~
[3,6].
The analogous Hspace construction, projecting the antiselfdual part
of the radiation field, is defined from the analytic extension of the complex ~
conjugate of (1.1).
Then the Hspace of an asymptotically flat Hspace is
simply flat i.e. complex Minkowski space.
These relations may be written
symbolically [3) 2
H
=
~
H; H H
~
0
H H.
The content of this introduction and §2 is drawn from many conversations with the relativity groups of the University of Pittsburgh and the Mathematical Institute, Oxford.
The identification of EIH with ~~: in particular is mainly
due to R. Penrose. The exact solution of §3 appears in [11] and the null geodesic equations for this solution have been solved in collaboration with E.T. Newman and W. Threatt. Responsibility for any errors and the vaguer speculations is my own.
§2.
THE IDENTIFICATION OF EIH
We show how the ~IH of an Hspace is identified with the ~~: of the spacetime M from which it arises with the aid of a series of figures showing
~IH,
asymptotic projective (dual) twistor spacePT* [4,7,8] and Hspace itself. Figure la shows EI: with a good cut Z.
Figure lb is the familiar cube
169
u
la:
~Jf~ with a good cut Z
lb:
~ ~ showing the real diagonal and a twister
j
line
w.
lying on Z
·
lc:
lP'3* with the compact A
holomorphic curve Z passing through W. Figure 1:
170
ld:
 z
The point z of r{space corresponding to the good cut Z
The relationships bett.1een cJ~
,
lPO* and r{space.
~r
z.
~

...
. ....  
2a:
The good cuts zi tipping up towards the generator r
2c:
The curves Z.~ through lying in a 2plane
~null
A typical zi meeting z 0 in w. with tangency at
2d:
The sequence of points z . on the geodesic y in
z
~
Hspace.
element Figure 2:
w.
2b:
geodesic in Hspace from the various points
of view.
171
picture of ICI: [4,7,8] showing the good cut Z ruled by geodesics or "twistor ~;;direction,
lines" in the
+
+
and indicating the real diagonal,
"'1;; = 1;;,
where
Figure lc is lPT*; the points of lPT* are the twistor lines of
a:IM meets lRIM.
A
figure lb and the good cut Z appears as a compact holomorphic curve logically s2
'
z,
topo
Finally, figure ld shows Hspace where the good cut Z
in lPT*.
represents a point z. Two points of
H are
null separated if the corresponding curves intersect
inlPT* [2,4,5] or equivalently, if the corresponding good cuts have a whole twistor line in common.
There will then be a point on the common twistor
line where the two good cuts are actually tangent.
A null geodesic y through
a point z 0 in Hspace now corresponds to the set of all good cuts which intersect the good cut Z
0
corresponding to z
same point of tangency, z.
in the same twistor line W, with the
0
Moving to points z.]. at larger and larger affine
distances from z 0 along y has the effect of tipping up the good cuts Zi more and more, as shown in figure 2.
Figure 2a shows the Zi tipping up and figure A
in W with tangency at z.
2 b shows one typical Zi meeting Z
0
0
of Zi on the diagonal in figure 2b corresponds to figure 2a. the curves
Z.].
The behaviour In figure 2c,
are shown meeting W0 and lying in a 2plane element defined by
the ~;;coordinate of ~. The tipped up good cuts approach the degenerate good cut which is simply the generator
+
r
A
of a:IM leading up from z.
Thus the "point
+
A
at infinity" on y may be canonically identified with the point z on ICIM. If the point z has coordinates (u
0
"'
,1;; ,1;; ) 0
whose corresponding good cuts pass through Z(z a ,1;; For fixed (u
0
o
"' ) ,1;; 0
= u
"'
,1;; ,1;; ), 0
0
0
then the points z of Hspace
z have
coordinates za satisfying (2.1)
0
(2.1) is the equation of a null hypersurface
"'
L
in
H
space (since the Hspace gradient Z a of Z at fixed I;; and I;; is null [2,4]).
' 172
Thus all the geodesics y which are regarded as having the same point at infinity ~ lie on the null hypersurface
r,
and this construction is seen to
mimic the "ideal points" construction of lRI~ [9]. The significant question now is whether EIH is suitably attached to Hspace, or equivalently whether the tipped up good cuts approach suitably smoothly to the degenerate good cut
r.
Since the tipped up good cuts will reach larger
and larger values of lui, this is a question about the behaviour of
o0 (u,~,~)
for large lui. Exactly what condition on 0° is necessary for asymptotic flatness of the Hspace is still unclear.
The example of §3 has
(2.2) and seems to be asymptotically flat in a reasonable sense. stronger than one might expect for a general M.
However, (2.2) is
In particular, there are
reasons for expecting that o0 will not even vanish for both large positive and large negative u in a single Bondi frame. out by B.D. Bramson [13].)
Typically though, there might be two Bondi frames
related by a supertranslation in one of which o positive u and in
th~
(This has also been pointed
0
behaves like (2.2) for large
other for large negative u.
A stronger condition than (2.2), which has arisen in conversation with
M. Ludvigsen,
E.T. Newmann and G.A.J. Sparling, is
"'
An(~.~) n~2
where an A Thus
A
n
n
u
n+l
(2.3)
0
r 2si~n
a
im n
so that the higher spherical harmonics in 0° fall off more rapidly as lui ~
00 •
Such a o 0 has a number of interesting properties (not all of which have been proved!):
173
a)
it is invariant under the Poincare subgroup of the BMS group,
b)
it is the condition for the Hspace to be regular at i+,
c)
if o 0 is used in the Kirchhoff integral [10] to obtain a spintwo zero
restmass field WABCD on Minkowski space then the null data of WABCD reproduces
o0
Equivalently, WABCD is regular at i+ or the twistor function defining
•
WABCD is analytic in the neighbourhood of the line at infinity laS in twistor space. It would be of some interest to know whether such a o 0 is likely to arise in a physically reasonable spacetime.
§3.
AN EXAMPLE OF AN HSPACE
It was first noticed by G.A.J. Sparling [12] that the good cut equation (1.1) was solvable with 0
"v
o (u,r;:,z:J
(3.1)
=
where f is a quartic polynomial, A is an arbitrary com?lex scale factor and a is an arbitrary nonzero real constant in~luded to make o 0 nonsingular on
+ RIM.
The specific case when f
=1
is dealt with at some length in [11].
This
example is sufficiently simple that the sorts of question raised in §§1 and 2 are tractable.
The solution may be summarised as follows: the good cut
"' is given by function Z(z a ,?;;,?;;)
. )2 (Z + 1a where z s "v
la(?;;,?;;)
z s
a a
(z + ia) "v
"v
"v
"v
2
(3.2)
(3. 3b)
la(?;;,?;;)
1
+ Xs
(3. 3a)
la(?;;,?;;)
l+ss
174
2
"v
(1,?;;,?;;,?;;?;;)
(3. 3c)
za s
(u,X,Y,v)
a
(3.3d)
1
X (Y,v,O,O)
(3.3e)
(u + ia)(v + ia) XY
(3.3f)
so that (u,X,Y,v) are the four complex parameters on which the solution depends. A
As
~
0 the solution becomes
z
z
which gives Minkowski space for the Hspace with the usual coordinates u
t  z
X
X
+ iy
The metric for nonzero ds 2
=
+ z
v
t
y
X 
(3.4) iy.
A is
2dudv 2dXdY ZA (Ydv(v+ia)dY) 2 •
(3.5)
~3
Since the metric is in KerrSchild form there is a canonical flat background and one expects the curvature to be algebraically special. curvature is type N, nonsingular away from
~
=
In fact, the
0 and has no singularities as
a function of A (being actually proportional to A).
Regarded as a linearised
solution on Minkowski space, (3.5) is an elementary state [7] corresponding to an imploding then exploding gravitational wave, everywhere nonsingular on the real Minkowski space.
In particular, it is asymptotically flat which
encourages the belief that it should still be asymptotically flat in the nonlinear theory. One problem with (3.2) is that one must take a square root to obtain Z explicitly and Z will not be a regular function on the sphere
~
~
Simple zeroes of the right hand side are avoided (i.e. no cuts).
=
~
unless
This is
175
reasonable, since a zero of the in a0
•
ri~ht
hand side corresponds to a singularity
Factorising (3.2) as
(Z + ia) 2 H 'V
H
'V
H H
(3 .6a)
z
+ ia +
! Hs
z
+ ia
iA!s
(3.6b) (3.6c) 'V
we see that simple zeroes of H or H are to be avoided.
It turns out that, 'V
with (u,X,Y,v) as in (3.4) and (t,x,y,z) all real, H and H are nonzero on the sphere
'V
s
=
s
provided
2 a •
(3.7)
With (3.7) holding therefore, there is a canonical real slice in the Hspace for which the good cuts are welldefined.
This can be thickened into the
complex to provide the real slab of §1 since
~
will be nonzero there, so
the metric and curvature are nonsingular. Note that, although the metric has no singularities as a function of A, the actual region of on A.
~4
for which the good cuts are welldefined does depend
If (3.7) is satisfied, the good cuts exist in a thickened neighbourhood
of the real slice in the canonical flat background.
Larger values of A will
have good cuts existing in a thickened neighbourhood of a translation of this real slice by an imaginary timelike vector. The next step is to solve the null geodesic equation for the metric (3.5). In currently unpublished work of Newman, Threatt and myself this has been done and remarkably it is possible to solve the equations completely although the metric has only one Killing vector.
There are two simplifying features
of the general Hspace which aid this calculation. of the good cut function Z at fixed
176
s
and
'V
s
Firstly, the gradient Z a
'
is null, so that at a point Z, 3
sweeps out the entire null cone as
~
and
'\,
~
vary.
This provides a scaling
for all null vectors and thus a preferred affine parameter on all null '\,
geodesics.
Secondly the gradient Z
•a
at fixed
where ; is the Hspace covariant derivative.
~
and
~
is actually geodesic:
a Thus a null geodesic z (T)
through the point za(O) in the direction of gab Z ~(zc(O),~.~) is a solution of the first order equations d a dT Z (T)
i.e. there are always four first integrals of the null geodesic equation in an Hspace and constants. Z
•a
when
is the preferred affine parameter, defined up to additive
T
One may now define "real" null directions as the ones given by
~ = ~.
This provides the antiholomorphic
for the definition of §1.
s2
of null directions
We now need to know whether, by going real affine
distances in real null directions from points on the real slice, we encounter the singularity or escape to infinity. initial position za(O)
= z~
This requires
and initial direction z~
~
as a function of T,
= gab
which is given by ~2
where
(Z
+ ia)
0
z0
Z(z
n
~
a 0
.~
~
~
~(za)
'\,
0
0
(3. 8)
)
0
0
i H 3/2
0
0
0
0
H
.~
0
~
0

0
(Tn)(TS)
H i ~ 3/2 0
s H
2
a

H(zo'~Z::o)
"'
a

H(zo,~o'~o)
177
Thus every null geodesic has two points a,8 in the complex Tplane where vanishes and the curvature is singular.
~
The question is whether, with the
reality conditions, these points can ever be on the real Taxis.
Unfort
unately this question is at present unresolved but the indications are that a and 8 are necessarily in the lower half Tplane provided (3.7) holds. this turns out to be true, then the Hspace is asymptotically flat.
If
Further,
if both singularities are in the lower half· Tplane we may take this as a definition for the Hspace to correspond to a positive frequency nonlinear graviton, [5].
It was conjectured in [11] that this solution would be
positive frequency in a reasonable sense, on the basis of the shear a0 being positive frequency in u and the linearised spintwo zerorestmass field corresponding to (3.5) being positive frequency in Minkowski space. In conclusion, we have an apparently rather coordinate dependent notion of asymptotic flatness for a left flat space arising from the Hspace construction.
"' The definition leans heavily on the good cut function Z(z a .~.~)
to define the real slice, real null directions and the affine parameter on null geodesics, which may be thought of as "memories" of the original real spacetime from which the Hspace arose.
It seems quite appropriate to use
this structure and it is gratifying that there is then an example other than flat space which (modulo a proof that Im a < 0, Im 8 < 0) satisfies the conditions of the definition.
REFERENCES 1.
E.T. Newman
in "General Relativity and Gravitation" ed. G. Shaviv and J. Rosen, Wiley, New York 1975.
2.
M. Ko, F..T. Newman and K.P. Tod in "Asymptotic Structure of Spacetime" ed. F.P. Esposito and L. Witten, Plenum Press, New York 1976.
178
3.
M. Ko, M. Ludvigsen, E.T. Newman and K.P. Tod, "The Theory of Hspace" submitted to Phys. Rep.
4.
R. 0. Hansen, E.T. Newman, R. Penrose and K.P. Tod "The Metric and Curvature Properties of Hspace", Proc.Roy.Soc. (Lond.) to appear.
l (1976) 31.
5.
R. Penrose,
G.R.G.
6.
M. Ludvigsen,
Ph.D. thesis, University of Pittsburgh, 1978.
7.
R. Penrose and M.A.H. MacCallum, Phys.Rep. 6C (1973) 242.
8.
E.J. Flaherty,
''Hermitian and Kahlerian Structures in Relativity" Lecture Notes on Physics 46 SpringerVerlag, Berlin 1976.
9.
R. Geroch, E.H. Kronheiner and R. Penrose, Proc. Roy.Soc. (Lond.) A327 (1972) 545
10.
E.T.Newman and R. Penrose, Proc.Roy.Soc. (Lond.) A305 (1968) 175
11.
G.A.J. Sparling and K.P. Tod, "An Example of an Hspace", J.Math.Phys. to appear.
12.
G.A.J. Sparling,
private communication.
13.
B.D. Bramson,
private communication.
K.P. Tod Department of Physics and Astronomy University of Pittsburgh, Pittsburgh, Pa 15260 U.S.A. and The Uathematical Institute Oxford England
179
J Isenberg and P B Yasskin Twistor description of nonselfdual YangMills fields §1.
INTRODUCTION
Recently there has been an increasing awareness of the physical importance of selfdual YangMills fields (e.g. their contribution to the Feynman path integrals of quantum chromodynamics)
[1]. ·The selfdual condition is, how
ever, a debilitating restriction for certain purposes:
(1)
fields on a real Minkowski space cannot be selfdual; (2) cannot solve the YM equations with sources.
Real YangMills Selfdual fields
Besides these restrictions,
the desire to obtain all of the contributions (no matter how small) to the quantum integrals leads one to consider nonselfdual YangMills fields. An important tool in the study of YM fields has been the Ward [2] correspondence between selfdual YM fields over complexMinkowski space, and certain holomorphic principal Gbundles over projective twistor space. The Ward correspondence, discussed extensively in this volume, has lead to the classification [3] and construction [4] of many selfdual YM fields. With an eye towards analogous classification and construction, we have modified the Ward correspondence and obtained a "twistorial representation" for the general nonselfdual YangMills gauge fields [5]. To motivate our modification we recall that the set of totally null complex planes in CM has two connected components: the aplanes, Pa.
the aplanes, Pa, and
The aplanes are spanned by antiselfdual bivectors and
parametrized by the projective twisters.
The aplanes are spanned by self
dual bivectors and parametrized by the projective dual twistors.
By using
Pa as the base space for its Gbundles, the Ward correspondence produces YM fields which are flat over each aplanea condition which is equivalent 180
to selfduality.
It is therefore clear that the modified correspondence for
nonselfdual fields cannot use bundles over P • a
Instead we use bundles
over L, the space of all complex null lines in CM.
The corresponding gauge
fields are flat over each such line, but this is of course no restriction; so we get the general nonselfdual fields. The properties of L, including its relationship to Pa and Pa,are discussed in §2.
The basic correspondence between bundles over L and gauge fields on
CM is stated as Theorem 1 in §3 and sketched as a geometric and algebraic construction in §4. Unlike selfdual gauge fields, general gauge fields do not automatically satisfy the currentless YM equations:
D*F
o.
(1.1)
While this face permits our correspondence (unlike Ward's) to handle fields with a current source, i.e.
D*F
*J
~
o,
(1.2)
it is also useful to know necessary and sufficient conditions on the bundle over L such that the corresponding YM field does indeed have vanishing current.
We state such conditions in the form of Theorem 2, in §3.
conditions are rather complicated and the proof is fairly long. of the purposes of this paper is to present this proof.
The
However, one
We do so in §5.
Other restrictions on the gauge fields can be realized as conditions on the bundle over L.
In reference 5 we state such conditions for reality,
flatness, selfduality, and holonomic commutativity. Note that the use of L instead of Pa (or Pa) entails a bonus:
Since the
null lines, in the guise of null geodesics, survive the passage to a general
181
curved spacetime, our correspondence should also survive and produce YangMills fields on a curved background.
The null planes do not generally
survive this passage and hence the standard Ward correspondence does not work for a general curved spacetime. §2.
LINE SPACE
Line Space, L, is defined to be the space of all null lines in conformally completed complex Minkowski space, CM.
We shall often want to discuss
fields which are not defined over all of CM, but only over a subset S C CM. So we define L(S) as the space of null lines which intersect S.
Similarly,
for r E S, we define L(r) as the space of null lines containing r.
These
definitions, and similar ones for Pa and P 6 , appear in Table 1. We assume that the reader is familiar with the spinor and twister parametrizations of the various spaces (our notation appears in Table 2.), and with the geometry of the Penrose correspondence [6] between these spaces (as exemplified by the definitions in Table 1).
Of particular importance is
the fact that every null line lies in a unique aplane and in a unique Bplane and is precisely their intersection.
This implies the existence of
the maps in the diagram,
where pa and
Ps
are projections and A is 1 1.
The map A is not onto.
In
fact an aplane, Z, and Bplane, W, intersect iff their twisters satisfy Z~
a
= 0,
in which case they intersect in a null line.
as the 5complex dimensional hypersurface, Z~a 182
= 0,
Hence, A imbeds L 3
3
within P a x P = CP xCP · B
a
S := subset of CM
= complex 1fold
p (S) == a
p (t) == unique aplane a containing £
£ := null line
P8 (s) := {null lines intersecting S}
p8(£) := unique 8plane containing £
W := 8plane
p (W) := {aplanes a intersecting W} = CP 2
I
I
L(S) := {null lines intersecting S}
£ : "' null line
1 p8 (W) :={null lines contained in W} = CP 2
p1 (Z) :={null lines a contained in Z) = CP 2
L L(r) := {null lines through r} = CP 1 X CP 1
W := totally null plane spanned by selfdual bivector complex 2fold
I {aplanes I intersecting S}
I
P8 (r) := { 8planes through r} = CP 1
Pe
P8 (z) := {8planes intersecting Z} = CP 2
through r} CP1
== {aplanes
p
Z == aplane
I Pa (r)
I
Z := totally null plane spanned by antiselfdual bivector complex 2fold
r := point
CM
a
r
AA'
wa A'
= 0]
[up to scales of r,A' and wA] [up to scales of r,A' and wA]
(vector, conjugate spinor, spinor) •••••• (ra,r,A''wA)
BL:
a
(vector, conj. spinor, vector, spinor) .• (p 'r,A' ,q ,wA)
(vector, conjugate spinor)
[up to scale of wA] a
a
a (vector, spinor) •••..••••••••••••••••••• (q ,wA)
[where zaw
[up to complex scale]
[up to scale of r,~]
a
)
[up to complex scale]
(p ,r,A,)
or complex 5tuple on each patch U(m] ••• xi:
(proj. twistor, proj. dual twistor) ••••. (Za,W) a
or complex 6tuple on each patch W[m] .•• (o,x )
E
a
(proj. twistor, proj. dual twistor) •.••• (Za,w )
or (spinor, conjugate spinor) ••••••.•••• (wA,IJ
projective dual twistor •••.•••.•••••.•••
or (spinor, conjugate spinor) ••••••••••• (nA,r,A,)
projective twistor •••••••••••.•••••••••• za
(pa + qa)/12 is the diagonal coordinate,
Ba xB8:
B8:
Ba :
L:
p axp a=
P a=
a
p:
and sa= (pa qa)/12 is the offdiagonal coordinate.]
[Here pa and qa are the coordinates on each factor, ra
or vector pair ••••••••.••••••••••••••••• (ra,sa)
a
vector pair • • • • • • • • • • • • • • • • • • • • • . • • • • • • • (p ,q )
=
CMxCM:
a
vector .................................. r
CM:
We also emphasize that L(r)
= CP 1
x
CP
1
is a compact complex manifold.
This is crucial to our construction which uses the fact that any holomorphic function on a compact complex manifold must be a constant. There are certain spaces which are irrelevant to the statement of the theorems, but are crucial to the proofs and also to the explicit statement of the correspondence.
These are the flag spaces£7]:
BL, Ba and Ba.
An
element of BL consists of a null line in CM and a point on that line.
Hence
there are projections nL CM BL L.
(2.2)
There are analogous definitions and projections for Ba and Ba.
The flag
spaces are related to each other by a diagram similar to (2.1).
The
coordinate representations of the various maps appear in Table 3. In the proof of Theorem 2, we need the commutative diagram,
CM b.
CM
1


nL
BL
1
1
t,.B
X CM~B
L A
(2.3)
n  P x Pa a x Ba a
We also use the transition functions for bundles over L(S), Pa(S) x Pa(S), BL{S) and Ba(S) x Ba(S).
So we let {W[m]} be a finite open cover of
Pa(S) x Pa(S) which is sufficiently refined so that any bundle over [mn] Pa(S) x Pa(S) can be specified by its transition functions, g , on W[m] n W[n].
We then define
u[m] := w[m] n L c L(S),
(2.4)
v[m] := n~ 1 (u[m]) c BL(S), and
(2.5)
x[m] := n 1 (w[m]) c Ba(S) x Ba(S).
(2.6) 18'5
....
Q
0\
AA' AA' (q ,wA) ~+ (wA,iq wA)
AA' AA' (p ,r,A,),.. (ip r,A,r,A,)
12
12
AA' i AA' i AA' (r ,r,A''wA) >+ (  r r,A''r,A''wA' r wA)
PARAMETERS AND COORDINATES FOR THE MAPS
a
~
a
~
or
a
+
L + p
A : u[m]
A
p
B
w[m]
X
(xr ) ~+
(O,x r )
A A' A A' (n ,r,A, ,wA,IJ ) ..... (n ,r,A, ,wA,IJ )
b.B: B1 ~ Ba x BB : (r ,r,A''wA) ~+ (r /t2,r,A''r /t2,wA)
a
6. : CM + CM x CM : (ra) ~+ (ra !12, ra /12)
, AA' AA' AA' AJ.' Ba x BB + pa x p B : I.P ,r,A' ,q ,wA),.. (ip r,A' ,r,A' ,wA,iq wA)
BB + P B
IIB
II
B + p a a
a
L
II
+
B1
IlL
TABLE 3:
Since diagram (2.3) commutes, we have
(2.7) Finally, we need coordinates on Pa(S) x Pa(S) which are adapted to L(S). Thus on each neighborhood, W[ml, we use coordinates (6,xE), E
= 1, ••• ,5,
where U[m] is the 6 = 0 surface and the (xE) are coordinates on U[ml.
§3.
THEOREMS
Using the language and notation of §2, we now state our theorems.
The first
of these describes the basic correspondence: Theorem 1 (Correspondence): Let G be a complex Lie subgroup of GL(n,C).
Let 5 be an open subset of the
conformal completion of CM whose intersection with each null line is either empty or connected and simply connected.
There is a 11 correspondence
between (1)
holomorphic principal Gbundles E5 over 5 with holomorphic connection A and curvature F; and
(2)
holomorphic principal Gbundles
~(S)
restriction EL(r) to L(r) for all r
£
over L(S) which have trivial S.
In §4, we describe how to build the line space bundle spacetime bundle E5 with its gauge field E5 , A and F,given EL(S)"
EL(S)' given the
A and F; and also how to construct
The proof [8], that these constructions are valid
and are inverses of each other, is analogous to that of Ward's theorem. Our other theorem states the conditions on EL(S) which are necessary and sufficient for the gauge field to satisfy the currentless YangMills equations:
187
Theorem 2 (Currentless): Let EL(S) be the bundle over L(S) which corresponds (according to Theorem 1) to the gauge field,
~
and FL, on the bundle ES over S.
Let gimn] be the
transition functions for. EL(S) appropriate to the cover U[m] of L(S).
Then
iff there exists a Gvalued function g[mn] on some neighborhood of U[mlnu[n] within each intersection, W[m] n W[n] cPa x P 6 , such that (a)
g[mn]l
L
= g[mn] L
and
~)
This theorem (proven in §5) says that the currentless YangMills equations are satisfied iff the bundle EL(S) has a third order extension from L(S) to some neighborhood of L(S) within Pa(S) x P 6 (s).
If the bundle can be
extended to all orders, then we get the much stronger restriction that the YangMills field also satisfies the holonomic commutativity condition [9]. In the
case of SU(2), this condition implies that the gauge field is self
dual, antiselfdual, or abelian. §4.
CONSTRUCTION OF THE CORRESPONDENCE IN THEOREM 1
The Correspondence Theorem implies that,using a given line space bundle EL(S)' one may construct a spacetime bundle ES with a connection FL; and vice versa.
~
and curvature
We now sketch how this is done.
The crucial fact in our construction, as well as Ward's, comes from compleX analysis:
Any holomorphic function from a compact, complex manifold into C
must be a constant.
We assume G is a complex Lie subgroup of GL(n.C).
any holomorphic cross section of a trivial 188
Then
principal Gbundle over a compact·
complex manifold is determined by its value at a single point.
Further, the
set of such cross sections may be identified with the fibre at that point. Given the bundle EL(S)' the point set of ES is constructed as follows: For each r e: S, let the fibre over r be the set 11 1 (r) of holomorphic cross sections of EL(r)' and let Es
= re:S u
compact and EL(r) is trivial, 11
1
1
11
(r).
Since L(r) = CP1 x CP 1 is
(r) may be identified with G.
We complete
the structure of ES by constructing a global trivialization in the form of a global cross section: along rrL: section.
Let the bundle EBL(S) be the pullback of EL(S)
BL(S)+L(S).
Since EBL(S) is trivial, choose a global cross
This section determines a holomorphic cross section of EL(r) for
each r in S and hence a global cross section of ES. The connection is constructed by first defining a transport T along null directions: holomorphic
Assume r and r' e: S lie on a null line !. c~css
section of EL(r')'
Each ~ e: 11l(r') is a
Since every holomorphic cross section
of EL(r) is determined by its value at a point, we define the transport of from r' to r along ! as the unique holomorphic cross section T
~
r~r
~(r)'
whose value at! e: L(r) n L(r') is
(Tr+r'~)(!) = ~(R.).
,
~
~
of
See figure 1.
This transport rule defines a covariant derivative provided it is linear in the differentiating direction. connection 1form If
~mn]
~
We check this by finding a formula for the
as follows:
are the transition functions of E1.(s) on U[m] n U[nl, then
g[mn]= g[mn] orr
L
L
(4.1)
L
[m] [ n] are the transition functions of EBL(S) on V n V •
Since EBL(S) is
trivial, its transition functions split: ~mn] = L
[m] ( [n])1
gL
gL
'
(4.2)
189
,
~,
I, II ' I
w
I I
...J
''
'
~
t,
''
''
)·'
''
') I
...JI ~I
Tt::
I I
l
'

I
C/)
'
I
''
~I
''
'
I I , I
'
I
"'...J
_J
u
(/)
_J
::JI I ' I
'~
,...
L.
~
I'
I
',
I
'
~
: ~; I :::I 't:: I I
C/)
w
'
'' ~
~
c
p..
~
z0
0
Cll
~
....
£:1 ~
...0
Cll
ri
H
~
g;j
~
fz<
t £:1
fz<
::t: £:1
:::::>
c
H
fz<
....
....
~ ~
0

•
''
r
''
' ' )"~ I I I
I I I
I
I
I I I I
t::\1
''
I I I
•
I
' ".J ,I
., I '
...
L.
J
I I
I I I I I
l
'
''
'~ I
I
''
I I .... I L.l I
' ' ,I ~
~
(.)
u (/)
where gL[m].~s defined on v[mJ.
(Note:
The choice of splitting corresponds to
the choice of global cross section of EBL(S) and to the choice of gauge on E8 .)
Using these gim],s we find that [10]
(4. 3)
is the desired 1form.
(We prove that
patch and independent of w and
~
(~)BB'
from (4.3) is independent of
by using
(4.4)
0, which in turn follows from (4.1).) The inverse constructionEL(S) from a given E8 with connection starts with the specification of the point set (of EL(S)): 1
let the fibre over i be the set n
(~
~again
For each i
E
L(S),
of covariantly constant cross
= U nl(R.). sections of Ei, the restriction of E8 to i; and let E L(S) iEL(S) 1 Since such a cross section is determined by its value at a point, n (R.) may
be identified with G.
We complete the structure of EL(S) by constructing a
local trivialization for each patch U[m] in the form of a map
~im]:n 1 (u[m]) ~
G:
.
holomorph~c
Pick a
on the line i.
A global trivialization of E8 provides a map ~ :E ~G. [m] [m] [m] function, r :U ~ S, such that the point r (~ lies
Each
8 8
~ E
1
n
which may be evaluated at r
(i) is a covariantly constant cross section of Ei
[m]
(~.
Thus we define
It follows that the transition functions are 
g~mn ] ( R.)
[m](i)
=
p exp{ 
f[ l
Jr n
A dx J.l } ,
(4.6)
(i) J.l
191
where "p" indicates that the exponential integral is pathordered. It can be shown [8] that these two constructions are inverses. §5.
PROOF OF THEOREM 2
Our proof of Theorem 2
i~
motivated by a correspondence between certain YM
fields on 5 x 5 and bundles over Pax P 8• statements are equivalent.
We show that a chain of five
The first and last appear in the theorem.
The
intervening conditions concern the extension of a bundle with connection from 5 to 5 x 5 and the extension of a bundle from BL(S) to Ba(S) x Ba(S).
(See diagram (2.3),)
The five conditions are as follows:
[I] [II]
(I.l) There exists a connection,
~
AA 1
= AAA1dr
1
+ BAA1dsAA , on Esxs
such that (a)
a * (~ =
~·
= (~)AAI;
and
(II.l)
DAA1
== ~ + AAA 1 and
(II.2)
vAA I
:=
AAAI la(S) (b)
or equivalently,
if we define
ar
a
;:;:o + BAA I '
(II. 3)
as
then we have (II.4)
(II.5) (Note:
DAA1 and VAA 1 are not really covariant derivatives and equations
(_II.4) and (II.5) are not really tensor equations.)
192
[III] There exists a connection, A= PAA 1 dp
AA'
+ QAA 1 dq AA' , on Esxs
such that (a)
!:J.*(!!:) ""~· or equivalently,
(III.l)
(b)
if we define a dAA' := ;;:Af + PAA' and ap
a
()AA'
:= ;;:Af
aq
then for all
(III.2)
+ QAA' •
rf'
and
(III. 3)
A
w
we have [10]
2 [dAi;'dBi;] = O(s ),
(III.4)
2 [()c..A'. ()t&B' ] = O(s ) , and
(III. 5)
[dAA I • ()BB' ] = 0(s 2 ).
(III. 6)
[IV] There exists a Gvalued function g Un] on some neighborhood of V Un] within each patch X[m] c Ba
x
B8 , and there exists a Gvalued function
g[mn] on some neighborhood of V[m] n V[n] within each intersection x[m] n x[nJ. such that (a)
g[m]IB
= g[ml,
(IV.l)
L
= 8 [m]
(g[n])1 + O(s4), and
(b)
8 [mn]
(c)
_a_ g[mn] = 0 and apA~;
_a__1 8 [mn] = aqwA
o.
(IV.2) (IV.3) (IV.4)
193
[V]
There exists a Gvalued function g [mn] on some neighborhood of U[m] nu ln] within each intersection, W[m] n W[n] CPa x P 6, such that
k = gimn]
(a)
g [mn]
and
(b)
g[mn]g[nk] = g[mk] + 0(64).
(V .1)
(V.2)
We now proceed to prove successively the equivalence of conditions [I] through [v ]. Transition [A): [I ] <>[II ] +)
We first assume the existence of a connection A= AAA,dr
AA'
+ BAA,ds
AA'
(A.l)
(A.2) Equation (II.l) implies that AAA'O
=
(~) AA''
(A. 3)
It can be shown that one can always pick a gauge, consistent with (A.3), in which the totally symmetric part of BAA'kB B'
B B' vanishes for all k;
i.e.
1 1'" k k
0.
(A.4)
In particular, (A.S)
Using (A.l) and (A.2) and this choice of gauge, [DAA''DRB'], [V AA''VBB'] AA' tole see that (II.4) and (II.S) and [DAA''VBB'] can be expanded ins 19~
imply DL * FL
0.
We also find that (A.6)
(A. 7)
(A.8) and (A.9) where (A.lO) (Notice that DL *FL = 0, (A.6), (A.7) and (A.9) actually follow from the weaker assumptions
(17,17)
(A.ll)
while DL * FL = 0, (A.6), (A.7) and (A.S) follow from
[17,17] = [D,D] + 0(s 1 ) and [D,V] = i*[D,D] + 0(s 2 ).) +)
Conversely, we assume
~
and FL satisfy DL * FL = 0.
(A.l2) In some gauge for
the bundle Esxs (specified by some global trivialization) we specify AA' AA' A= AAA 1 dr + BAA 1 ds by giving AAA' and BAA' as the quadratic polynomials AAA' = AAA'O + AAA'lBB's and
BB'
+ AAA'2BB'CC's
BB' CC' s '
(A.l3)
BAA'
= BAA'O
+ BAA'lBB's
BB'
+ BAA'2BB'CC's
BB' CC' s '
(A.l4)
with coefficients given by (A.3), (A.S), (A.6), (A.7), (A.8) and (A.9). Equation (A.3) guarantees that (II.l) is satisfied. as in (II.2) and (II.3).
We define DAA' and VAA'
Then, if we substitute (A.l3) and (A.l4) into the
exapnsions for [DAA''DBB' ], [VAA''VBB'] and [DAA''VBB']' we find that (II.4) and (II.S) are satisfied. Transition [B]: [II] The same
~>[III]
connection~
appears in conditions [II] and [III].
The coefficients
PAA' and QAA' may be related to AAA' and BAA' by using the coordinate transformation
r
AA'
1
=  (p
12
AA' + q AA' ),
p
AA'

1
12
(r
AA'
+ s AA' ),
AA'
AA'
(B.l) q
AA'
1
=
1:2
(r
 s
).
Thus we find that
AAA'
BAA'
1
(PAA' + QAA,)'
1 PAA' =  (AAA' +BAA,)'
/:2 (PAA'  QAA,)'
1 QAA' =  (AAA, BAA,).
=
12 1
=
1:2
(B.2)
12
Consequently, we have dAA' = _! (DAA' + VAA,)'
12
(B. 3)
The first of equations (B.2) shows th4t equation (II.l) is equivalent to equation (III.l).
196
Using equations (B.3) we relate the commutators:
(B.4)
Z[VAA 1 'VBB 1
)
=
[dAA 1 'dBB 1
+ [oAA 1 ' 0BB 1
)
)
(B.S)  [dAA 1 ' 0BB 1 )

[oAA 1 ,dBB 1
),
(B.6)
(B. 7)
uAA I • ur BB I ) = (oAA I • DBB I ) + [v AA I • VBB I )
2[ r
AA I • VBB I )  [v AA I • DBB I ) •
(B.8)
[D
2 [d
r
AA 1 'uBB 1
= [D
)
AA 1
D '
BB 1
) 
[V
V
AA 1 ' BB 1
)
(B.9)
It follows that equation (II.4) and (II.S) are equivalent to (B.lO)
(B.ll) [dAAI' \BI)
=
2
(B.l2)
O(s ).
Equation (B.l2) is just (III.6).
It remains to prove that (B.lO) is equi
valent to (III.4) and that (B.ll) is equivalent to (III.S).
To see this,
we write (B.l3)
197
(This expansion can be carried out for any antisymmetric tensor.)
Using
CC'DD'
(B.lS)
i EM'BB' along with the definition of
"*"• (A.lO), we find that
(B.l6)
(B.l7) Consequently, (B.lO) and (III.4) are both equivalent to 2
(B.l8)
= O(s ),
'l'A'B'
while (B.ll) and (III.S) are both equivalent to
AAB
= O(s 2 ).
Transition [C): [III]
(B.l9) ~>[IV]
We first assume the existence of a connection
+
)
~
= PM,dpM' +
QM,dq
M'
satisfying conditions (III.l)(III.6).
We expand
PM' and QM' in the coordinate sAA', obtaining
(C.l)
PM' BlBi QM'
=
k:O QM'kBlBi···BkBk s
BkBk ••• s
(C.2)
The functions g[m] and g[mnJ, as yet unknown, may also be expanded: (C.3)
198
g
[mn]
(C.4)
We specify g[m] and g[mn] by specifying the expansion coefficients in (C.3) and (C.4).
To guarantee that (1V.l) is satisfied, we require that
[m]
gL
(C.S)
•
[m] [m] Expressions for glAA'' gzAA'BB''
[m] g3Az;;BB'CC'
[m] and g 3wA'BB'CC' may be read out
of the expansion of the equations
(C.6)
a wA' aq
[m]
(C. 7).
g
provided certain consistency conditions are met.
These conditions are
guaranteed by (i11.1), (111.4), (111.5) and (111.6).
We now have all the
[m] [m) ~ ~ components of g 3AA'BB'CC' except g 3 w~z;;~w~z;;~w~z;;~' where z;; and w are spinors
such that z;;~, z;;
AI
A
.I.
= 1 and wA w
= 1.
For later purposes, it is necessary to
require that
o.
(C.S)
The higher order terms of g[m)are arbitrary provided that (C.3) converges. Hence we set
[m]
gkAlAi·. ·~Air_
0,
for k
~
4.
(C.lO)
This completes the definition of g[mJ.
199
To guarantee that (IV.2) is satisfied, we define [mn] gkA1Ai·· for k
·~Ak
= 0,1,2,3,
(C.ll)
where h
[n]
=
(g
[n] 1 ) •
To second order, (IV.3) and (IV.4)
are satisfied by virtue of (C.6), (C.7), and (C.ll). We then inductively define the higher order terms in g[mn] in a manner which guarantees that (IV.3) and (IV.4) are ·satisfied to successively higher orders.
In the first step of the induction, consistency is guaranteed
[mn] is left undetermined. kA1 Ai···~A~ It is chosen to guarantee consistency at the next order. The induction
by (C.S).
At each order, one component of g
completes [11] the definition of g[mn] and shows that (IV.3) and (IV.4) are satisfied. +
)
Conversely, we assume the existence cf g[m] and g[mn] satisfying
(JV.l)(JV.4).
We define a function
flm](r,s,~,w),on
each patch
X[m] c Ba x Be, as the quadratic polynomial in sBB' given by the zeroth, first and second order terms in the power series expansion of r~[m))1 'Q
_a_ [m] A~g • ap
substitute (IV.2) into (IV.3) to obtain (C.l2)
and hence ( g [m) )1
a ap
~g
[m]
( g [n])1
a ap
~g
[n] + 0(s3).
(C.l3)
Since flm]and fln]have been defined to be the quadratic parts of each side of (C.l3), it follows that
200
(C.l4) Thus there is a global function
fA(r,s,~,w)
on Ba x Ba, consistently defined
by  fA[m] (r,s,~,w),
fA(r,s,~,w)
where
(r,s,~,w)
(C.l5)
X(m).
E
Since g[m] is homogeneous of degree zero in
~ and
w,
the functions
(g[m])la A~ g[ml, flml, and hence fA, are homogeneous of degree zero in w ap and of degree one in
~.
Since fA is globally defined, there exist functions
PAA,(r,s) such that (C.l6) Therefore, for any
(r,s,~,w)
E
X(m),
a [m] + 0(s3). ( g [m])1 ~g ap
(C.l7)
Similarly, there exist functions QAA,(r,s) such that (C.l8) Using (C.l7) and (C.l8) we compute
....! a (g[mlls=o>· 1:2 I s=O = 1 arw~
(C.l9)
Then combining (IV.l) with (C.l9) we obtain
(C.20) Since w and
~
are arbitrary, this proves that (III.l) is satisfied. 201
Next, we define dAA' and oAA' according to (III.2) and (III.J).
Using
(C.l7) we compute
_a_[ ( [m])1_a_ [m]] + ( ,[m])1 (a [m]) ( [m))l(_a_ [m]) AI; g B~; g g AI; g g Bl; g ap ap ap ap
2
O(s ).
(C.21)
Similarly, using (C.l8) we compute (C.22) Finally, using both (C.l7) and (C.l8) we compute
_a wB' [(g[m))1 a AI; g[m)) aq ap
[m]) ( [m])l(_a_ [m])+O( 2) ( g [m])l(_ll_ wB' g g Al;g s aq ap
= 0(s 2 ). For each
1;
and w, the computations (C.21), (C.22), and (C.23) are true for
some patch X[ml.
Hence, (III.4), (III.S), and (III.6) are satisfied.
Transition [D): [IV] +)
(C.23)
~>[V)
We first assume the existence of Gvalued functions g[mnl, on the
patches of Pax P6 , satisfying conditions (V.l) and (V.2). projection n:Ba
202
X
BB
~ pa
X
Using the
PB, we pull the functions g[mn] back to
Ba x Ba and define g
[mn]
= n*g [mn]
[mn] g o n •
(D.l)
It follows from (D.l) and the coordinate description of IT that, for any spinors >.
A
and P
A'
, we have
(D.2) That is, g
[mn]
is invariant under translations
and along the aplane {q
AA'
,wA).
along the aplane (p
AA'
.~A')
Therefore, (IV.3) and (IV.4) follow from
(D.3)
0 and
Since the diagram (2.3) commutes, we see that condition (V.l) implies (IV.l): [mn]
gL
[mn]
o ITL = gL
•
(D.4)
To prove condition (IV.2) we first show that the g[mn] satisfy the cocycle condition to order s 3 : (D.5) l"e prove (D.5) by expanding both g(mn] and g[mn] in power series:
g
[mn] (D.6)
[mn]
g
[mn]
Using the chain rule, we may express the gkA A' [mn]
gk
~~·
in terms of the
1 1' '"1Ck
•
For example,
a
[mn]
glAA'
=
[mn]
g AA'
as
Is=O
ag[mn] ~ ag[mn) ~ AA'I + a6 as s=O ax I: as AA'I s=O
+
[mn]a6 =gl as
w 1s=O
ag[mn] I: o ax I: ~1 ax as s=O
Condition (D.S) then follows directly from the assumption (V.2). We finally show that the third order splitting condition (IV.2) follows from the third order cocycle condition (D.S) and from the analytic properties Recall that the bundle EB
over BL is trivial, and so any set of L
transition
functions
g~mn]
for EB
must split.
Therefore, we have
L
[mn] = sJmnJ go L
[m] (g n)1.
go
o
(D.S)
Starting with (D.S) and using the fact that the cohomology class
(D.9) vanishes, one can show that the first, second and third order terms in (D.5) [mn] [mn] d [mn] lit i th th t imply that glAA' • g2AA'BB'' an g3AA'BB'CC' sp • n e sense a condition (IV.2) is satisfied. + )
Conversely, we assume the existence of functions g[m] and g(mn]
satisfying (IV.l)(IV.4).
While there is no map from Pa x Pa to Ba x Ba
which might be used to pull g[mn] back to Pa x Pa, we see that the conditions 204
(IV.3) and (IV.4) say that the g[mn] satisfy (D.2).
Hence, we may define
g[mn] by choosing any pair of points p AA' and qAA'
such that p
A
the aplane (n .~A') and q
AA'
is on the Splane
g
[mn]
(p
AA'
,q
AA'
A'
(wN~
AA'
is on
) , and then setting
'~A''wA).
(D.lO)
This g[mn], of course, satisfies (D.l). We now consider g[mn] restricted to L C P for any point in L must meet. AA'
we may pick p
q
AA'
The aplane and Splane S Therefore, in evaluating g[mn] from (D.lO), a
x P .
Using (D.lO), (IV.2), (IV.l) and a relation
similar to (D.lO) which relates gimn] and gimn], we obtain condition (V.l): [mn]l L  g [mn]l B
(D.ll)
g
L
We finally prove that the g[mn] satisfy the third order cocycle condition (Y.2).
Condition (IV.2} immediately implies (D.S).
We again expand g[mn]
and g[mn] as in (D.6) and use the chain rule to relate the coefficients as However, we now want the [mn] in terms of the [mn] gk gkAlAi •• ·~Ak and not vice versa. We can use the chain rule (D.7) only because of the in (D. 7}.
freedom to chooses
AA'
=
0 whenever
o = 0.
While we cannot solve (D.7) for
the g[mn) in terms of only the g[mn) we can obtain the gk[mn) in k kA 1 Ai···~~· terms of the g [mn) and the g [mn) for j < k. Then condition (V.2 )
~Ai·· ·~Ak
j
follows order by order from (D.5). Acknowledgments This
wor~
crucial Thls
t .~
was done in collaboration with Paul Green, whose insight was the forillulation of our theorems.
~"ork ~•as
supported in part by the National Science Foundation under
Grant PHY 76·20029.
205
References 1
R. Jackiw, S. Coleman,
Rev. Mod. Phys. 49, (1977), 681706; in Erice Lectures 1977, ed. by A. Zichichi, (Academic Press, NY, to be published); contribution to this volume.
R. Jackiw,
s.
2
R.
Ward,
Phys. Lett. 61A, (1977) 8182.
3
M. F. Atiyah and R. S. Ward,
Commun. Math. Phys. 55, (1977), 117124. M. F. Atiyah, N. J. Hitchin and I. M. Singer, Proc. Nat. Acad. Sci. USA 74, (1977). 26622663.
4
M. F. Atiyah, N. J. Hitchin, V. G. Drinfeld and Y. I. Manin, Phys. Lett. 65A, (1978), 185187; N. H. Christ, E. J. Weinberg, N. K. Stanton, "General SelfDual YangMills Solutions," 1978, Columbia Univ. preprint CUTP119.
5
J. Isenberg, P. B. Yasskin and P. E. Witten,
6
R. Penrose, R. 0. Hansen and E. T. Newman,
s.
Green, Phys. Lett. 78B(l978), 462464; Phys. Lett. 77B(l978), 394398, in Quantum Gravity, ed. by C. J. Isham, R. Penrose, and D. W. Sciama, (Clarendon, Oxford, 1975), 268407; Gen. Rel. Grav. 6 (1975), 361385.
7
The space Ba is the same as the space F 12 discussed by R. Wells, contribution to this volume.
8
Available upon request.
9
See Theorem 4 in Isenberg, Yasskin and Green, ref. 5.
10
[m]
If we regard gL
0 in ~A' and wA' then
a
;;;AI : = ar 11
A w
a
AA'
as functions of (r
a A A' a w := w ~ ;;:};' ar ~
ar
~A''wA),
, •
a ;;p; := ar
homogeneous of degree ~
A'
a
;:AI • ar
and
;:AI ar
Note that we have not yet verified that this series converges. J ISJ:NBERG and P B YASSKIN Center for Theoretical Physics Department of Physics and Astronomy University of Maryland College Park, Maryland 20742
206
E Witten Some comments on the recent twistor space constructions
In this talk I would like to comment on certain features of the HorrocksBarthAtiyahHitchinManinDrinfeld (HBAHMD) construction [1] of bundles in ~F
3
and selfdual YangMills solutions.
(This construction will also be
described at this conference by Robin Hartshorne.) One purpose is to explain in a simple language why this construction works, and incidentally to translate the twistor space construction into Minkowski space language.
However, the main motivation for the effort described here
is that I feel that the construction as usually presented has a number of features that are too special for broad applicability in physics. In particular, I would like to deemphasize those aspects of the construction that are inherently global.
Although the treatment of the global
Euclidean space problem is very powerful and beautiful, we must remember that physics takes place in Minkowski space and that global considerations are not usually in the forefront in physics.
Also, the quantum YangMills field
satisfies not the selfdual equation F 0.
'\,
jJV
=
F
jJV
but the second order equation
If we hope to eventually make applications to quantum field
theory, we must probably learn to think about this latter equation in Minkowski space.
And this means not a global problem but a local or
initial value problem.
In particular, it would be very exciting to find an
analogue of the l!BAHMD construction for the Minkowski space equation DJJF
jJV
With t.:::s in mind. I will be describing the HBAHMD construction in a way that :Ls suited f:or
3
consideration of the local problemthat is, the
problE::m of finJinr, selfdual YangMills fields that are well defined, not
207
0.
throughout Euclidean space, but just in a small open set thereof. corresponds, in cc:JP 3 ,
(This
to a bundle defined not on all of CClP 3 but just on a
neighborhood of a line.) First, let us review Ward's construction of the selfdual YangMills solutions [2].
Ward considers (cf. his paper in this volume) the space of
all lefthanded null two planes a in four dimensional complex Minkowski space CClM. bundle on CClP
Given a selfdual gauge field !n CCIM, Ward introduces a vector 3
as follows.
An element of this bundle is a pair (a,~)
where a is a null two plane in CClM and
~
is a (Lorentz) scalar field (in the
fundamental representation of the gauge group) that is covariantly constant on a. We may say that
~
is defined only on a.
But a statement that is more
suitable for generalizations is to say that (rather, throughout as much of
CC~
~
is defined throughout CClM
as we are working with) but is defined
only modulo the addi.tion of a field vanishing on a.
According to this
definition, in other words, the fiber of Ward's bundle over a given two plane a is the space of all scalar fields space of all fields that vanish on a. (a,~)
~
with D
ll
~
=0
on a, modulo the
The bundle E consisting of all pairs
has, as Ward showed, a natural holomorphic structure, from which the
original selfdual YangMills gauge field can be reconstructed. Now, HBAHMD tell us that this bundle E can be obtained as follows. introduce vector spaces A, B, and C, and for each point Za introduce. linear maps f(Z):A
~
B, g(Z):B
~C.
We
in ctlP 3 we
For fixed Z, f and g are
linear maps, and also they depend linearly on the homogeneous coordinates Za of CC:JP 3 • A
208
f(Z)
Also, for fixed Z, g(Z)f(Z) B~ C
= 0.
The whole picture is
(1)
and the bundle E arises as the kernel of g modulo the image of f.
In other
words, for fixed Z, the fiber of E at Z is the kernel of g modulo the image of f. I would like to explain what A,B,C,f, and g are in the construction of reference [1], and why they are that way. As a first approximation,, let us imagine that A, B, and C each are the space of all scalar fields (in the fundamental representation of the gauge group) and let us try to define maps f and g such that E will be the kernel of g divided by the image of f. We will define f, corresponding to a given Z or to a given two plane a, to be multiplication by a suitably chosen function which vanishes on a.
Then
im f will consist precisely of scalar fields that vanish on a. It is a little harder to define g.
We will define g so that the kernel of
g consists of all scalar fields that are covariantly constant on a.
Then
ker g/im f will be exactly the space of all covariantly constant fields on a. The above is a good approximation to what we want, but it suffers from a basic deficiency.
We are trying to construct selfdual gauge fields.
dual gauge field is
~ompletely
A self
determined by certain arbitrary functions of
three variables, which one can choose, for instance, to be the values of the gauge field on the initial value hypersurface t
= 0.
However, the spaces
A, B, and C, as defined above, are spaces of arbitrary functions of four variables.
At best, it is extremely redundant to describe the selfdual
gauge fields, which really depend only on functions of three variables, in terms of spaces A, B, and C that involve arbitrary functions of four variables. The
pJ.·~
principle.
blem
l.S
not just a problem of redundancy; it is a problem of
If our goal is to find a construction along the lines of (1)
that could be used, at least in principle, to solve the initial value problem,
then A, B, C, f, and g should be spaces, and maps, that are known explicitly once the initial data are given.
I think that this is a reasonable property
to insist on in a construction like [1]. To arrange that A, B, and C are spaces of functions of three variables, not four, and that they (and the maps f and g between them) are such as to be known explicitly once the initial data are given, we should choose A, B, and C to be spaces, not of arbitrary scalarfunctions, but of solutions of some auxiliary equations in the background selfdual gauge field.
Then A, B,
and C will be spaces of certain arbitrary functions of three variablesthe initial data of the auxiliary equations. To proceed further, I must be more specific, and tell you what function that vanishes on a we will use in defining f. The function we use will be not a scalar but a spinor, (2)
where (cA,d
A' ) are constants, the choice of which depends on a, and where
xAA' are the coordinates of complex Minkowski space. Now, you will recognize that
~A
=0
defines a two dimensional surface
which is in fact a lefthanded null two planea surface corresponding to a twistor or element of aP 3 • homogeneous coordinates of
A' can be regarded as the
(In fact, cA and d ~p
3
.)
We will sometimes identify a two plane a with the spinor (2) that vanishes on it. I will also need the spinors of opposite helicity ~
~A'
210
(3)
It is an essential, and not completely obvious, fact that the spinor spaces
~ and ~ are dual to each other.
In fact, we can define a bilinear form (4)
This form is obviously Lorentz invariant.
It is not obvious that it is
conformally invariant, but this is, in fact, true. Now, let us define the spaces A, B, and C. A will be a space of unprimed spinors uA that satisfy a certain equation that will be described later. B will be a space of scalar fields
~
that satisfy a certain equation that
also will be described later. C will be the space of the unprimed spinors wA which satisfy the Dirac equation
o.
(5)
Now let us define our maps f and g. choice of two plane in
~IM
Remember, f and g will depend on the
or, equivalently, on the choice of the spinor
~A
which vanishes on the two plane. We define f:A
~
B by (6)
Then the image off consists of functions that vanish on the two To define g I must first state what equation the scalar fields satisfy. (13) for
plane~ ~
A
=0.
in B
This equation will be given in a rather implicit form (but see eq. ~
mor£ explicit version) and at first sight looks rather peculiar.
The condition to be satisfied by
~
is that its covariant derivative has
an ,;xpansion
211
0) where ~A'a' a=l, ,,, ,4 are our four spinors of primed type defined in a equation (3), and where wA, a=l, ,,, ,4 are required to be solutions of the
Dirac equation, that is, elements of C. Now we can define our map 4
l.\
g(ljl) $
'\a
(1j1,1jl
g(~):B ~C.
We define
a
)wA
a=l We must show that the kernel of g consists precisely of those scalar functions that are covariantly constant on the two surface on which 1jl vanishes. To argue this, I will proceed in a noncovariant way. 1, d 2 ' = 0, that is 1!1A = xAl,.
Consider the special
Any case could be mapped
onto this one by a Lorentz transformation. Also, I will choose a basis for the primed spinors "' ljiA,: ~(1) 1'
1,
"(2)
o,
1jl 1'
"(1)
2'
0
"(2) 1jl 2'
1
1jl
(9}
"(3) 1jl
A'
xlA'
"(4) 1jl
A'
x2A''
In this basis the formula DAA,$
, .....
l.~A'
a
wA
a
means that
a
(10)
212
Also,
(~.~) is nonzer~ with our choice of ~ and in this basi~ only for ~(l).
Therefore
g(~)~. with out definitions, is zero if and only if wA(l)
=
0.
In
this case (11)
so that DAl'~
=0
if xBl'
= 0.
(12)
This is the desired result, stating that to be in the kernel of be covariantly constant on the two plane on which
~
g(~).
~must
vanishes.
As one can see by differentiating equation (7) and symmetrizing on the primed indices, antisymmetrizing on theunprimedones (and using the fact that wA satisfies the Dirac equation while
~A'
satisfies the twistor equation
3AA'~B' + 3AB'~A' = O) to get a consistency condition, fields ~ that satisfy (7) exist only if the background gauge field with respect to which D
~
is
defined is selfdual; it is here that selfduality enters into the construction. Now let us tie up the loose ends in the definitions of A, B, and C.
C
was already defined as the space of solutions of the Dirac equation (5).
B
we define as the space of all scalar fields whose covariant derivatives can be expanded in the form (7), and A we define as the space of spinors uA such that, for
any~
A
of type (2),
uA~
A
is an element of B.
are rather indirect and are not immediately transparent.
These definitions With some thought
one can see that a more explicit and equivalent way to define B is to say that B consists of the space of all scalar fields AA' 2 a+ X bAA' +X c
~
that can be written (13)
213
where a, bAA'' and call satisfy the covariant Laplace equation, D~ D~s
= 0.
However, a given element of B can be written in the form (13) in many ways, as a result of which this expansion
~s
awkward to use.
The general element
of A can be defined in a somewhat similar way. Since we have shown that the kernel of g consists of scalar functions covariantly constant on a, and the image of f consists of scalar functions vanishing on a, it may seem quite plausible.that the kernel of g modulo the image of f is precisely the fiber of Ward's bundle E.
What still must be
shown is that B is large enough that any desired covariantly constant value on a is assumed by some element of B, and that A is large enough relative to B that the equivalence classes, the kernel of g modulo the image of f, depend only on the value on a.
These facts can be established by relatively
elementary arguments analogous to those above. Rather than continuing in this vein, let us shift here to a more formal line of argument.
Manin and Drinfeld [1] have given a very elegant deriva
tion of the construction under discussion here using the Kazul complex, which is an exact sequence of sheaves. product of the cotangent bundle of
Let nk be the kth antisymmetric tensor ~p
3
.
Then there is an exact sequence
of sheaves
depending on the arbitrary choice of a point S in ~P 3 represents· the complex numbers sitting at S).
(and where ~Is
Following Manin and Drinfeld,
one tensors (14) with E(1) and writes down the standard long exact sequences of cohomology groups, in order to obtain information about E. one finds that if one defines
211.1
In this way
® o2 (1))
A
H1 (E
B
H1 (E ®g)
(15)
and f and gas the natural maps induced from (14), then E is in fact ker g/im f. The point I wish to make here is that one can easily see that global properties are not needed in this argument.
If instead of all of CP 3 one
is working with a neighborhood of a line in CP 3 , still goes through.
the ManinDrinfeld argument
(Some of the arguments for why certain cohomology groups
vanish are modified, but the net conclusion is the same.) a line in
~p
3
A neighborhood of
corresponds to a small open set in complex Minkowski space,
and thus it corresponds to the problem of a selfdual solution that is defined only in a small open set.
Thus, the ManinDrinfeld derivation shows
that the construction that we are discussing here is valid for this local problem as well as for the global problem. In fact, from the Minkowski space point of view suggested in this talk, the construction is "obviously" right in the local problem, while its validity in the global case is a
d~licate
fact.
The strategy described in
the first three paragraphs after equation (1), and followed in this talk, will certainly work if A, B, and C are defined to be large enough spaces of functions.
But if they are too small it could happen that ker g/im f would
be not the whole fiber of E, but only a subspace.
It is most clear that this
does not happen if A, B, and C are defined without restriction as spaces of functions of four variables, as suggested in the comments just after equation (1).
With the definitions actually given in this paper, A, B, and
C being defined in terms of functions of three variables, it is slightly 215
delicate but still relatively easy to see that the construction works.
In
the global problem treated in reference one, where the elements of A, B, and Care required not only to satisfy conditions (5), (6), and (7), but also to be global elements of the cohomology classes, these spaces become finite dimensional.
In this case it is a quite delicate fact that the spaces A, B,
and C are still large enough that the construction works. Thus, this construction is also valid in a local form.
As I mentioned at
the outset of this talk, the reason that I think that this fact has some significance is related to the fact that the second order Minkowski space YangMills equation, D~F
~v
0, is quite probably the equation we must come
to grips with if we hope for applications to quantum field theory.
These
equations are hyperbolic; one can think of them in terms of a local problem or an initial value problem, but there is for them no global problem analogous 4
to the global selfdual problem on S •
The fact that the HBAHMD construction
makes sense in a local version encourages one to hope that an analogue of this construction may exist for the second order YangMills equations.
The
discovery of such an analogue would be very exciting. A suitable starting point in a search for such a construction might be a construction recently found for the second order YangMills equations by Yasskin, Isenberg, and Green [3] and by Witten [4].
This construction has
been presented at this conference in talks by Yasskin and by Isenberg. I would like to conclude with several technical remarks.
In the local
problem, A, B, and C are infinite dimensional vector spaces.
The use of
infinite dimensional spaces should not alarm or surprise us.
It is analogous
to the fact that in the inverse scattering appraoch to the sinegordon equation, the sinegordon field is reconstructed form linear maps among
216
certain infinite dimensional vector spaces, the spaces of solutions of the auxiliary Schroedinger equation [5]. It should be noted that in the local problem, f is not injective, although g is still surjective.
If one wants to find a sequence of spaces and maps
exact except at one stage, the nonexactness corresponding to E, then it is necessary to introduce a fourth vector space; roughly speaking, the fourth space is the kernel of f.
Also, it should be noted that certain simplifying
features of this construction that appear in the global case are missing in the local case:
A is not dual to C, B does not have a natural bilinear form,
and g is not the adjoint of f (in the global problem these properties appear if E is symplectic or orthogonal). The last point to be made here concerns the question of the choice of gauge groups or of the rank of vector bundles.
Most discussion so far has
concerned SU(2) gauge fields or rank two bundles.
For applications to
quantum field theory, however, the often made assumption that SU(2) is the simplest theory is probably an error. Rather, it was shown by 't Hooft in a very important paper [6] that has not received all the
at~ention
it deserves that YangMills theory with an
SU(N) gauge group is, as a quantum field theory, simplest in the limit N
+ ~.
I believe that the most realistic goal in this field is to try to understand the large N limit proposed by 't Hooft. It is very striking that the HBAHMD construction also simplifies (for fixed value of the Chern class) as N
+ ~O].
This convergence of the region
in which the quantum field theory is known to simplify and the region in which the mathematics of classical solutions simplifies is very intriguing, and perhaps of great significance.
It may well turn out to be, in the long
217
run, the most significant aspect of the construction, if one has physical applications in mind. I would conjecture that a generalization of the HBAHMD construction to the second order equations, if it exists, will be something that is uselessly awkward for any finite N, and is tractable only at N
= ~.
References 1
G. Horrocks, unpublished; W. Barth, unpublished; M. Atiyah, N. Hitchin, Yu. I. Manin, and V. G. Drinfeld, Phys. Lett, 65A, 185(1978). Yu. I. Manin and V. G. Drinfeld, Comm. Math. Phys., to be published.
2
R. Ward,
3
J. Isenberg, P. B. Yasskin, and P. S. Green, "Non SelfDual YangMills
Phys. Lett. 61A(l977) 81. Fields," to be published in Phys. Lett. B.
"An interpretation of Classical YangMills Theory," to be published in Phys. Lett. B.
4
E. Witten,
5
Some additional comments on this analogy have been made by L. Fadde'ev (private communication).
6
G. 't Hooft,
Nucl. Phys. B72(1974) 461.
7
M. Atiyah,
private communication.
Acknowledgment I would like to thank D. Kazhdan and D. Mumford for discussions. E WITTEN Lyman Physics Laboratories Department of Physics Harvard University Cambridge, UA 02138
218
J Harnad, L Vinet and S Shnider Solution to YangMills equations on M4 invariant under subgroups of 0(4,2) The purpose of this paper is to present, in outline, a systematic exposition of what is known about solutions to the YangMills equations on
4
compactified Minkowski space M which are invariant under a subgroup of the conformal group 0(4,2). 0(4) x 0(2) on
4
If we identify M with U(2) then the action of
M4 is up to a covering map equivalent to the action of
SU(2)L x SU(2)R x U(l) where the subscript L or R indicates multiplication on the left or right respectively.
We give a complete list of all solutions
invariant under a)
SU(2)L x SU(2)R
b)
SU(2\
x
U(l)
and all self dual solutions invariant under c)
SU(2)L (or SU(2)R)
d)
(SU(2)L x SU(2)R) diag.
First some notation: Let
x 2+ix) E
Herm(2)
xOxl v
=
(I+ ih)(I ih)l E U(2)
219
22 where u 0+u 5 sign.
2222 u 1+u 2+u3+u 4
=1
and the vector
(ua)E~
6
is determined up to
Then h ~ v imbeds lR 4 in the compact manifold U(2) as an open dense
set in such a way that the pseudo metric det(v 1dv) is conformally related to the standard Minkowski pseudo metric det(dh). (x~)
Cartesian coordinates
The correspondence between
and group coordinates (ui) is given by
0, 1, 2, 3. The coordinates (ui) identify the twofold cover !.'14 of Further if u 4
= cos~
and u 5
system on SU(2) ~ s3, t
x0 , r =
~~+x~+x~
t+r
tan (~~)
~
= cos~,
~
M4 with
SU(2) x U{l),
being part of a spherical coordinate
an angular coordinate on U(l)
"'= sl,
and
, then
2
(1)
tr =tan (~;~).
A basis for the leftinvariant forms is given by (2)
wi = 2(u4dui uidu 4 + ujd~ ukduj) where (i,j,k) is a cyclic permutation of (1,2,3).
Let {oj} be the Pauli matrices and {Xj = oj/2i} the
corresponding basis for the Lie algebra su(2). Any su(2) vector potential (local connection form) on ~ can be written (3)
The associated field (curvature) is
n
220
= dw
+ 21 [w, w] •
(4)
The sourcefree YangMills equations are (5}
where the * operation on two forms is defined by the pseudometric
If the curvature form is selfdual or antiselfdual ±iO
(6}
the YangMills equations are a consequence of the Bianchi identities.
We
are looking for solutions satisfying a condition of invariance up to gauge transformations of the form (L*w) = AdP(g,u)l(w ) + P(g,u)ld P(g,u), g u u u for gEG, one of the groups listed above.
Two theorems which are very helpful
in reducing the problem are given below. Theorem 1 Let (g,u) ~ f (u) be the left action of SU(2) on ~4 • g
Given any
principle SU(2) bundle E over ~4 • n:E+ ~4 • on which there exists a group action (g,b) ~ ~ (b) such that n~ (b) = f (n(b)) there is a bundle isog g g "'4 morphism T:E+ M x SU(2). Furthermore by the appropriate choice of T (appropriate gauge) we can write Tt~ T1 (u,h) = (f (u),h), (u,h) g
g
~4
EM
x
SU(2).
This means the local expression for the vectorpotential is globally defined on ~4 and the gauge invariance function P can be chosen to be P(g,u)
= e. 221
Theorem 2 '1.4 Let (g,u) + f (u) be the conjugation of SU(2) on M (which is the induced g
action of the covering group of S0(3) acting on class of SU(2) bundles n:E
+
~4 ); then in each equivalence
~4 (indexed by the integers [5]) there exist
exactly two inequivalent group actions satisfying n(~ g (b)) = f g (n(b)).
We
can always find a trivialisation • of E over (U(l)  {e}) x SU(2) such that either
•~ ,l(u,h) g
or
1
'V
't
g
(u,h)
=
(f (u) ,h)
=
(f (u) ,gh).
g
g
In this case not all bundles with S0(3) action are trivial so the vector potential is not necessarily globally defined. reduced to either P(g,u) : e or P(g,u) results and references.
= g.
The gauge function P can be
We can now give
a table of
(See following page.)
In order to study cases 1, 2 of the table we look at the full SU(2)L invariant equations
(Sa)
and (8b)
where~=
~
=
222
i
AkXi and B
f~; and B
= Bi Xi.
In case 1 the SU(2)R invariance implies
= O, therefore there is only one equation
SU(2)L
(SU(2)L x SU(2)R)diag (equivalent to 0(3) invariance)
3)
4)
b)
a)
=
f *(I) g
(I)
=
*
f( g,g )(&)
= Adg 1w
f*(g,g)w
(I)
(I)
=
f *w g
(I)
* f(g,h)(l)
b)
=
* 1 f(g,h)w"" Ad h w
a)
Invariance equation
=0 Bi(ljl)
w
i
B (!p,ljl)
Bi(ljl)
+ Bi(!p,ljl)dljl)
= Xi(Ai(!p,ljl)d~
Aj(iP,'jl)
i
A~(ljl)
i i Aj, B constant
Ai j
Bi = 0
A~(ljl) = f(ljl) c5~
Form of ansatz in equation (3)
Spherically symmetric static solutions on M4 are studied in [7][9)[11][5].
U(2)L
2)
§
SU(2\ X SU(2)R (equivalent to 0(4) invariance)
1)
Invariance group
Invariance conditions and known solutions
All solutions [5]
All (anti) selfdual solutions, some nonselfdual [5][7][9][ll] §
All (anti) selfdual solutions [5) some nonself dual solutions [5][6]
All solutions [3)[4][5][6]
Solutions pure gauge
All solutions [3)[8][10][12][13]
What is Known
t f" + 2f(lf)(l2f)
=
o.
Luscher [10] and Schechter [13] have given a complete analysis of this class. The equations can be solved explicitly using elliptic functions. constant solutions f = 0, 1 are pure gauge but the solution f
The 1
=2
is non
trivial and in another gauge had already been found by deAlfaro, Fubini and Furlan [3].
The self dual solution was analyzed by Rebbi [12].
In case 2, A1 and Bi are constant.
We define X
See also [2],
= (X~) = (~~) and equations
(8) become 1
X(l + trX)
=
XB where B2
+ 3(detX)ZI  x2 + !(~XB 4
 B2X)
0 (8 I)
(trX)B
= EBiBi.
Solving these equations for X, Band then looking for
those A such that X
= At A
and which also satisfy the original equations, one
finds Theorem 3 Up to constant gauge transformation the U(2)L invariant solutions to (8) are
I •. A= a®y
with a, BE su(2)®¢, y
B = B
where a. B = 0 and B2 = 4 either a 2 = 0 or y 2 = 0 II.
B
where R E S0(3, ¢)
224
0
E
su(2) *
l.U.
A
IV.
B
A = AI 1
=0
1
= 2•
B arbitrary
A= 0
see Harnad et al. [5].
Howe and Tucker [6] and Tucker and Zakrewski [14]
give partial results to this and the SU(2)L case (case 3). In case 3 we look at the (anti) selfdual equations
(9)
However if the vectors {Ai} span a one dimensional subspace at each point, then by first gauge transforming
to
B = 0 the full equations (8) become
linear and are solvable. Next we transform the equation by A = j
l
 j
e± 2 iljl
B = ~e±2iljl
and introduce the complex variable w
If
= e+2iljl •
Then
{li} spans a two dimensional subspace we can solve (10) immediately;
otherwise we introduce the matrix
(Y~)
=
(li.lj)
(11)
Let g be such that Y=g.I+C
225
where C is a constant complex symmetric matrix with trace C
0.
Then g
satisfies 2 (~) dw
= det Y = g 3 + ag + b
(12)
where a, bare integration parameters for (11).
We solve for gin terms of
elliptic functions then determine Y by specifying the remaining parameters in C, which we assume to be in a canonical form (using complex conformal transformations).
Finally, we look for
l
such that Y = llt and
l
satisfies (10).
Theorem 4 The SU(2)L invariant solutions to equation (6) are as follows: 1)
If
{l1 ,l2 ,l3 } y
(I
j
span a !dimensional space at each point and B = 0 11 e su(2)®G: Yj scalars
2)
If
{l1 ,l2 ,l3 }
span a 2dimensional space at each point then up to a
constant gauge transformation and a cyclic permutation
l 2 .. a l 3 2 (11cos6+asine) ~ = 2(llsine+acose) where e is a complex parameter and 11,6 E su(2)®G:. 3)
"'
If the vectors Ai are linearly independent, then three classes of solutions to eq. (5.6) exist, depending upon whether the matrix Y has two (and hence three) eigenvectors of nonzero length, only one,or none at all.
Denoting by
l
the matrix whose columns are
to a gauge transformation, B vanishes and 226
l1, l2,
and
l3,
then up
(i)
If Y has three eigenvectors of nonzero length
where p = bds (b(ww0 )
m)
bns (b(ww0 )
m)
r = bcs (b(ww0 )
m)
q
ds, ns and cs are JacobiGlaisher functions, b, m, w0 E S0(3,~)
R E
(ii)
~
and
are arbitrary.
If Y has only one eigenvector of nonzero length, iq pq 0
where
p
= """"":"'..~a=~ sh[a(w  w )]
0 b sh[a(w  wo>]
q = · a
r =
b coth[a(w w0 )] a
with a, b, w0 E (iii)
~.
ab ol 0 and R E
S0(3,~)
I f Y has no eigenvectors of nonzero length,
l
=
[M i:
iq pq ir
~j
R
where
227
1 = =~
p
w  wo
with a, b, w0 E C,
a~
0 andRE S0(3,C).
In case 4 with invariance conditions a) we write
w
=Au
"'
w0 + BW + C uXW + D u (u.w) where u
"'
"' "'
"' "' "'
"'
=
(u 1 , u 2 , u 3 ) and w
...
and A, B, C, Dare scalar functions of~. ~· 1 2 2 2 When (B  2> + E /~ ~ 0 where E = u + 2u 2 C, we can find a gauge in which B
4
1
When E
= 2"
E~ A  +i Esin~
~
"'
0 we can write the equations as D
(13)
epsin~ and the last equation becomes
Let E
e
Let ~ = 1jl~
n
2
P~
n
=
e
2P
= 1jl~ 2
and we have
2P
which has a general solution
228
PCS, n) "' ln ( E2
cr•F(~)G(n) a>G'
= F' (~)G' (n)
sin
(F(~)G(n)) 2
2
(~n).
In order to have regularity at the spatial origin F=G and both of period n. F(~)
f
(tan~)
~
0
(~=n)
we must have
Write G(n)
g(tanll)
and transform to cartesian coordinates; the generating function E becomes
E(r,t)
f'(r+t)g'(tr) (f{r+t)g(tr))
2 • r
2
The resulting solutions are regular on M4 provided f=g and is strictly monotonic, however the solutions will not be regular on the compactified space ::4 M• The S0(3) invariant selfdual solutions in Euclidean space were first studied by Witten [15] and his analysis carried over to Minkowski space is equivalent to ours under a suitable change of gauge and basis.
For some
::4 particular spherically symmetric static solutions on M see Hsu and Mac [7],
Ju [9] and Protogenov [11]. References 1
w.
2
J. Cervero, L. Jacobs,
3
v.
Phys, Rev,, D 16, 3609, (1977).
Bernreuther,
c.
Hohl,
Phys. Lett., 69.!, 351, (U77},
4
deAlfaro, S. Fubini, G. Furlan, Phys. Lett., 65 B, 163, (1976). and CERN preprint TH2397 Oct., 1977. Phys. Lett., 76 B, 589, (1978). J. Harnad, L. Vinet,
5
J. Harnad, L. Vinet, S. Shnider,
CRM preprint 792 (submitted to J. Math. Phys.), part~two in preparation. 229
6
P. Howe, R. Tucker,
CERN preprint TH2421 (to appear in Nucl. Phys.)
7
J. P. Hsu and E. Mac,
J. Math. Phys., 18, 100(1977).
8
R. Jackiw, C. Nohl, C. Rebbi,
"Classical and semiclassical solutions of the YangMills Theory" in Boal and Kamal, Particles and Fields, Plenum Press 1978.
9
r.
Ju.'
Phys. Rev., D 17, 1637 (1978).
10
M. Luscher,
Phys. Lett., 70 B, 321 (1977).
11
A. P. Protogenov,
Phys. Lett., 67 B, 62 (1977).
12
c.
Rebbi,
Phys. Rev., D 17. 483 (1978).
13
B. Schechter,
Phys. Rev., D 16. 3015 (1977).
14
R. Tucker,
15
E. Witten,
w.
Zakrewski,
CERN preprint TH 2462. Phys. Rev. Lett., 38, 121 (1977).
Acknowledgment This research was supported by NRC grants •
.T HARNAD and L VINET
Centre des Recherches Mathematiques Universite de Montreal Montreal, Canada and S SHNIDER Department of Mathematics ~1cGill University Uontreal, Canada
230
P Green Integrality of the Coulomb charge in the line space formalism We will consider the Coulomb field as the curvature of a connection on a trivial complex line bundle whose base space is the complement of the world line of a point charge, taken to be the taxis in Minkowski space.
Having
chosen an appropriate background trivialization, we may identify the connection with the oneform qdt/r.
Here r 2
= x2 +
y 2 + z 2 , and q is an
appropriate multiple of the electric charge. As long as we confine our attention to the field on real Minkowski space,
"'q may
take on any real value.
What we will show in the present note is that
if we complexify the field and require that the resulting complex field be obtainable, as in [1], from a bundle over the space of null lines in its domain, it follows that "' q must be an integer.
"'
In order to interpret qdt/r as a complex field, we take as its domain, V
= {x,y,z,t,r)lr 2 = x 2 +
y 2 + z2
~ 0}.
Since r has exactly two possible
values for each choice of (x,y,z,t), Vis a double covering of
Q = {(x,y,z,t)lx 2 + y 2 + z 2 = 0}.
c4
 Q where
This is necessary in order to make sense
of r. The hypothesis that the field is obtainable from a bundle over the space of null lines in V is equivalent to the existence of a globally covariantly constant section of the bundle over every null line. line is
x
=
cover.~{)
t
=
A, y
=
z
=
0.
This line intersects Q only in the origin.
=
It
in V by two copies on which r = A and r = A respectively.
Restricting our attention to the copy on which r
"'qdt/r
Let us consider the
'"qdA/A.
A, we may write
The parallel translation law may be described as follows: 231
Let
~
be a section which is constant in the background trivialization.
Then the parallel translation of may be written
~
~(Y(O))
exp({qdA/A)~(Y(l)),
along the curve Y from Y(O) to Y(l)
If there exists a globally covariantly
constant section over the line, then parallel translation around the line must be independent of path.
In particular it must be the case that
~
exp(JqdA/A) = 1 whenever Y is a closed path. y
It follows that, for every ~
~
closed path Y, /qdA/A is an integral multiple of 2ni. y
Since JqdA/A is y
~
evidently 2niq when Y is any closed curve which encircles the origin once, ~
we deduce that q is an integer. Although the general null line is more complicated (see remark 2 below), ~
it can be verified that the integrality of q is a sufficient condition for global integrability on each null line. Remarks 1.
~
Until the precise relation between q and the electric charge is clarified, ~
the physical implications of the integrality of q are not clear. 2.
The space of null lines of V is a nonHausdorff manifold owing to the fact that a generic line in C4 meets Q in two points and is nontrivially covered by its preimage in V, while there is an exceptional set of lines which meet Q in a single point or (even more exceptionally) not at all. Each of these exceptional lines is trivially covered and its preimage consists of two null lines of V instead of one.
The geometry of this
line space is very interesting and I hope to return to it in another paper. 3.
4
The necessity of passing to a twofold covering of C  Q is somewhat obnoxious, particularly since it raises the possibility of passing to a still higher covering,which would rob the result of all its force. may, of course, be desirable in view of Remark 1 above.)
232
(This
Two points can
be raised in this connection.
The first is that higher coverings would
be of a different geometrical nature from the one that defines V; only the extremely exceptional lines which miss Q altogether would be trivially covered.
The second point is that the introduction of V is less ad hoc
if ?ne reflects on the fact that line bundles on V correspond in a natural way to a class of 0(2,C) bundles on
c4
 Q.
This suggests that
one should take the background of the Coulomb connection to be a nontrivial 0(2,C) bundle (which becomes trivial when lifted to V) rather than a trivial line bundle. References 1
Yasskin, Isenberg, Green,
NonSelfDual Gauge Fields. Phys. Letters B, to appear.
GREEN Department of Mathematics University of Maryland College Park, Maryland 20742
PAUL
233
N H Christ Unified weak and electromagnetic interactions, baryon number nonconservation and the AtiyahSinger theorem There are two quite distinct classes of physical phenomena which are currently believed to be described by a YangMills theory: interactions (nuclear forces) where the
quan~a
1.
The strong
of the gauge field and the
isolated "charges" to which it couples are imagined to be unobservable (confinement).
2.
Weak and electromagnetic interactions.
Here the photon
is an honesttogoodness gauge particle and the other gauge quanta whose exchanges mediate the weak interactions (e.g., Bdecay or neutrino scattering) are very massive and as yet unobserved.
However, it is widely expected that
examples of these particles will be discovered within the next decade. In this talk I would like first to describe the simplest model of such a YangMills theory of weak and electromagnetic interactions, the WeinbergSalam model [1].
Second, following 't Hooft [2], I will discuss the relation
ship between the Pontryagin index, the Adler anomaly and the nonconservation of baryon number.
Finally I will show how the AtiyahSinger theorem imple
ments this nonconservation of baryon number in a semiclassical, Euclideanspace calculation. The WeinbergSalam model is built upon the gauge group SU(2) x U(l).
In
i
addition to the four gauge fields {A~}l~i~J and B~ corresponding to the generators ·of SU(2) x U(l) this model contains a doubtlet of scalar fields {~a(x)}
a= 1 • 2
belonging to the fundamental representation of SU(2), a number
NL ~ 12 of similar doublets of lefthanded Fermi fields {ljiL~ (x)} 1
{ljiRi(x)}l~i~N
and L , unaffected by SU(2) transforma
R In addition to the electromagnetic field A (x) which is a linear ~
combination of the gauge fields A3 and B only the Fermi fields ).1
)J
wai
correspond
to known particlesthe "known" quarks and leptons. The gauge fields obey the following equations [3]
(la)
(lb)
L R where yi and yi are simple fractions chosen so that the quarks and leptons have the proper electric charges, {(; i )Ba} 1~~3 are the three 2x2 Pauli and we have used a vector notation to represent the three SU(2)
matrices indices
i
of AJ.I.
If all terms nonlinear in the fields are dropped from Eq.(l) i
i
then only derivatives of the fields AJ.I and BJ.I appear.
This would imply that
the corresponding four quanta are massless particlesa physically untenable situation.
Consequently, the WeinbergSalam model postulates a selfinter
action of the field ~~ so arranged that in the configuration of minimum energy the field ~a(x) is a nonvanishing constant.
~ 1 (x) = 0, ~ 2 (x)
Thus, if we choose
= v/1:2, substitute into Eq. (1) and then drop all nonlinear
terms, the resulting equations are 1 _tV+g'e B"1112  ~[ga v 3 A
(2)
where e 3 is a unit vector in the three direction.
The nonderivative terms
on the right hand side of Eq. (2) give the corresponding quantum mechanical ~35
particles a mass:
1
2
1
the quanta of A and A have mass zglvl the quanta of
1122 1v I while those of g'A3+gB have mass zeroa gA3+g'B have mass 1fg'+g'' physically acceptable spectrum if glvl~50 GeV. As observed by 't Hooft, the conservation law for the baryon number
current in this model contains an Adler anomaly [4]: (3)
where NB is the number of doublets of lefthanded quark fields, Fi
lJV
is the
field strength tensor (4)
and ~i
lJV
its dual.
The baryon number jBll is given by (5)
the sum extending over all right and lefthanded quark fields.
The quarks
carry baryon number ±1/3 and are the only particles with nonzero baryon number.
The Adler anomaly (i.e., the right hand side of Eq. (3)) is a
quantum mechanical effect and reflects the slow convergence of sums over largemomentum quantum states.
The anomaly is not present in the classical
field theory where baryon number is exactly conserved. Since the time component j~(x) is the baryon number density, the increase in total baryon number QB between the times ti and tf is QB(tf)  QB(ti) = Jtf t.
l.
(6)
where a surface integral at spacial infinity has been dropped. limit tf + +.,., ti +
co
Thus in the
the change in baryon number is NB v/3 where v is the
usual Pontryagin index for the YangMills field.
Consequently, if gauge
field configurations with nonvanishing Pontryagin index play a role in the quantum mechanics, one expects a violation of baryon number in the WeinbergSalam model and a resulant instability of nuclear matter! Of course such an observation is interesting only to the extent that one can actually calculate the size of the effect. scheme has been developed by 't Hooft.
The following computational
One begins with a functional integral
expression for a Green's function containing 2n Fermi fields: G(x 1 , ••• ,x
2 n)~ J d[A~]
x exp{iJ
d [Bv] d
[~a]
d
[~)
d
[~]
[LYM+~+LF]d 4 x}
(7)
x[~il (xl) ••• ~in (xn)~in+l (xn+l) ••• ~i2n (x2n)] Here JLYMd 3x, JL8 d 3x, and JLFd 3x are the Lagrangians for the YangMills fields, the·scalar field and the Fermi fields respectively. Perhaps the least familiar part of the "Feynman path integral" (7) is the integration over the Fermi fields
~i'
If we treat the gauge and scalar
fields as fixed, then we can expand each Fermi field
~(x)
in terms of the
eigenstates Xi(x) of the Dirac operator
v = iyll[a,..
A~ • l (ly 5 )i ~ yLB (1y )i ~ y~ (l+y )] 4 'IJ 2 i'IJ 5 2 i'IJ 5
(8)
"'  h'ij'~' ,T  h ij'~'
which appear.; in LF: (9)
237
The Dirac matrices (1 ± y 5 )/2 are projection operators onto right/left
=
fLFd4x where
A~
q~'s
In terms of the
handed states.
J
= L A~q~t
~V~d 4 X
the Fermi action becomes
(10)
q~
~
is the eigenvalue of the operator (8) corresponding to
functional integral over the Fermi field integral over the coefficients
q~
and
t
q~.
~(x)
X~·
The
is then replaced by a multiple
These are assumed to be anti
commuting algebraic quantities and all integrals computed according to the rules [s]
fdq~ = fdq~ fdq~ q~
=
(lla)
=0
fdq~
Note higher powers of
q! q~
1.
(llb)
automatically vanish since
q~ q~
of their anticommutation. In order to compute low energy,baryon number violating contributions to the Green's function (7) one begins in Euclidean space and continues to physical Minkowski space only after the amplitude has been completely evaluated.
The Euclidean integral over Ai, Band ~a is done by a saddle
point method.
If the scalar field were absent the extrema of the action,
ILYMd 4x, with Pontryagin index v
1 would be configurations with
B~
Ai a oneinstanton solution with position z and scale parameter p. ~
=
0 and
The
inclusion of the scalar field ~a with the selfinteraction required by the WeinbergSalam model spoils these extrema but in a very simple way.
If the
gauge coupling g is very small compared to the self interactions of
~a. then
the one instanton solution is extremal with respect to all variations except that of p.
238
The action is an increasing function of p and only for very small
2 2 instantons, p << 1/lvl, does it become the usual 8w /g •
There are no
exact classical solutions to the coupled YangMills and scalar field equations. This p dependence of the action is not a serious obstacleit is possible to perform onedimensional integrals which are not Gaussian.
In fact it is
an enormous assistance because configurations of large, overlapping instantons (which obscure the application of these methods in the case of the strong interactions)are now highly suppressed. The baryon number conservation equation (3) implies that for a YangMills oneinstanton configuration like that above, the resulting Green's function should violate baryon number by NB/3 units.
Similarly a configuration with
Pontryagin index v must contribute to a Green's function violating baryon number by NBv/3 units.
This relationship between the Pontryagin index v and
the above Green's function calculation is directly implied by the AtiyahSinger index theorem.
In order to understand this we must be somewhat more
specific about the nature of the Fermi field we continue to Euclidean space.
~(x)
appearing in Eq. (10) when
In Minkowski space that part of the action
which connects the SU(2) gauge field and Fermi fields in the two dimensional representation of SU(2),
NL
~ L
i=l
I4
d x ~L i iy (3 i _&+t2 T•A )~L i' ' 'II 'II 'II '
(12)
can consistently be written in terms of a lefthanded doublet hermitian conjugate.
a
~L,i
and its
However, in Euclidean space, the Dirac matrix Y'll
connects lefthanded fields with the hermitian conjugates of righthanded fields so that the
~
in Eq. (12) must be a full fourcomponent Fermi field
containing both leftand righthanded fields.
However, since the Weinberg
Salam model always contains an even number of lefthanded Fermi fields with 2J9
couplings to the SU(2) gauge field this is always possible.
One must simply
use for one half these
~R,i
Fermions the charge conjugate fields
in which the roles of particles and antiparticles
c
t = Y 2 ~L,i
are reversed and which
are righthanded [6]. We can now ask for the consequences of the AtiyahSinger theorem.
In the
approximation that g is small compared to the scalar selfinteraction, the B and 1.1
~
terms in Eq. (10) can be ignored.
Thus the action becomes a sum
over NL/2 identical terms
(13)
where each
~i
is now a fourcomponent Dirac field made up of left and right
handed parts as described above. ..& ...
The AtiyahSinger theorem implies that the
.....
operator iy (a i 2 A •T) has v vanishing eigenvalues. '1.1 '1.1 '1.1
The corresponding
eigenstates are lefthanded if v is positive and righthanded if v is negative. for the i
Let {q . • } 1 l., ....
th
field
~i'
~
R.
~v
1 <
l..
2
<
be the coefficients of these v eignestates
lN
L'
Since the Euclidean Dirac matrix y'll
connects left and righthanded states, positive v implies that the qi,R. are t
coefficients of lefthanded eigenstates while qi,R. appears in the righthanded part of
~i'
Because of the integration rules (lla) these variables t 4 qi,R. and qi,R.' being absent from !Ld x, must appear in the explicit fields ~i (x1 ) ••• ~. (x 2 ) in the integrand of Eq. (7) if the contribution to 1 2n 1 n t
Thus integration over the qi,R. and qi,R. requires the presence of vNB/2 lefthanded fields
~L,i
with baryon number
1/3 and vNB/2 righthanded fields ~ct also with baryon number 1/3. R,i
Since
each such operator represents the annihilation of a particle carrying baryon
240
number 1/3, the corresponding Green's function describes a process in which a total baryon number of NBv/3 disappearsexactly as predicted by Eq. 's (3) t
and (6).
(The integration over the remaining q's and q 's is sufficiently
symmetrical that no additional baryon number violation results.) Because of the presence of the precise number of eigenstates with vanishing eigenvalue implied by the AtiyahSinger theorem, this Euclidean space calculation agrees nicely with the baryon number violation predicted by Eq. (6).
However, our Euclidean space calculation also determines the
size of the effect. v = 1.
For NB
=6
the minimal baryon number change is 2 when
4
The factor exp[ /LYMd x] for a one Euclidean instanton configuration
2'1T sin 26 ] where in the WeinbergSalam model becomes exp [  a w sin26 = g' 2 /(g 2 + g' 2) = 1/4 and the fine structure constant a= 1/137,
w
giving an overall factor of e215 •
This factor is sufficiently small that
very few deuterons have decayed since the universe was created!
Of course
we might speculate that the complete theory of "weak and electromagnetic" phenomena might require a larger gauge group in which instantons of smaller action can occur, allowing predictions of the type just discussed but of greater physical interest. References 1
S. Weinberg, A. Salam,
Phys. Rev. Lett. 19, 1264(1967); in Elementary Particle Physics, N. Svartholm, ed. (Almquist and Wiksells, Stockholm, 1968) p. 367.
2
G. 't Hooft,
Phys. Rev. Lett. 37, 8(1976); Phys. Rev. Dl4, 3432 (1976).
3
We use the notation of Bjorken and Drell, Relativistic Quantum Fields (mcGrawHill, New York, 1965).
4
S. L. Adler,
Phys. Rev. 177, 2426(1969).
5
F. A. Berezin,
The Method of Second Quantization (Academic Press, New York, 1966).
6
The author thanks Savas Dimopolous for pointing this out to him.
7
A. Schwartz,
Phys. Lett. 67B, 172(1977);
M. F. Atiyah, N. J. Kitchen and I. M. Singer, Proc. Natl. Acad. Sci. USA 74, 2662 (1977). L. S. Brown, R. D. Carlitz and C. Lee, Phys. Rev. D 16, 417(1977). Acknowledgment This research was supported in part by the United States Department of Energy.
N H CHRIST Department of Physics Columbia University New York, New York 10027
242