In theoretical physics, the BRST formalism, or BRST quantization (where the BRST refers to the last names of Carlo Becchi, Alain Rouet, Raymond Stora and Igor Tyutin) denotes a relatively rigorous mathematical approach to quantizing a field theory with a gauge symmetry. Quantization rules in earlier quantum field theory (QFT) frameworks resembled "prescriptions" or "heuristics" more than proofs, especially in non-abelian QFT, where the use of "ghost fields" with superficially bizarre properties is almost unavoidable for technical reasons related to renormalization and anomaly cancellation.
The BRST global supersymmetry introduced in the mid-1970s was quickly understood to rationalize the introduction of these Faddeev–Popov ghosts and their exclusion from "physical" asymptotic states when performing QFT calculations. Crucially, this symmetry of the path integral is preserved in loop order, and thus prevents introduction of counterterms which might spoil renormalizability of gauge theories. Work by other authors a few years later related the BRST operator to the existence of a rigorous alternative to path integrals when quantizing a gauge theory.
Only in the late 1980s, when QFT was reformulated in fiber bundle language for application to problems in the topology of low-dimensional manifolds (topological quantum field theory), did it become apparent that the BRST "transformation" is fundamentally geometrical in character. In this light, "BRST quantization" becomes more than an alternate way to arrive at anomaly-cancelling ghosts. It is a different perspective on what the ghost fields represent, why the Faddeev–Popov method works, and how it is related to the use of Hamiltonian mechanics to construct a perturbative framework. The relationship between gauge invariance and "BRST invariance" forces the choice of a Hamiltonian system whose states are composed of "particles" according to the rules familiar from the canonical quantization formalism. This esoteric consistency condition therefore comes quite close to explaining how quanta and fermions arise in physics to begin with.
In certain cases, notably gravity and supergravity, BRST must be superseded by a more general formalism, the Batalin–Vilkovisky formalism.
BRST quantization is a differential geometric approach to performing consistent, anomaly-free perturbative calculations in a non-abelian gauge theory. The analytical form of the BRST "transformation" and its relevance to renormalization and anomaly cancellation were described by Carlo Maria Becchi, Alain Rouet, and Raymond Stora in a series of papers culminating in the 1976 "Renormalization of gauge theories". The equivalent transformation and many of its properties were independently discovered by Igor Viktorovich Tyutin. Its significance for rigorous canonical quantization of a Yang–Mills theory and its correct application to the Fock space of instantaneous field configurations were elucidated by Taichiro Kugo and Izumi Ojima. Later work by many authors, notably Thomas Schücker and Edward Witten, has clarified the geometric significance of the BRST operator and related fields and emphasized its importance to topological quantum field theory and string theory.
In the BRST approach, one selects a perturbation-friendly gauge fixing procedure for the action principle of a gauge theory using the differential geometry of the gauge bundle on which the field theory lives. One then quantizes the theory to obtain a Hamiltonian system in the interaction picture in such a way that the "unphysical" fields introduced by the gauge fixing procedure resolve gauge anomalies without appearing in the asymptotic states of the theory. The result is a set of Feynman rules for use in a Dyson series perturbative expansion of the S-matrix which guarantee that it is unitary and renormalizable at each loop order—in short, a coherent approximation technique for making physical predictions about the results of scattering experiments.
This is related to a supersymplectic manifold where pure operators are graded by integral ghost numbers and we have a BRST cohomology.
From a practical perspective, a quantum field theory consists of an action principle and a set of procedures for performing perturbative calculations. There are other kinds of "sanity checks" that can be performed on a quantum field theory to determine whether it fits qualitative phenomena such as quark confinement and asymptotic freedom. However, most of the predictive successes of quantum field theory, from quantum electrodynamics to the present day, have been quantified by matching S-matrix calculations against the results of scattering experiments.
In the early days of QFT, one would have had to say that the quantization and renormalization prescriptions were as much part of the model as the Lagrangian density, especially when they relied on the powerful but mathematically ill-defined path integral formalism. It quickly became clear that QED was almost "magical" in its relative tractability, and that most of the ways that one might imagine extending it would not produce rational calculations. However, one class of field theories remained promising: gauge theories, in which the objects in the theory represent equivalence classes of physically indistinguishable field configurations, any two of which are related by a gauge transformation. This generalizes the QED idea of a local change of phase to a more complicated Lie group.
QED itself is a gauge theory, as is general relativity, although the latter has proven resistant to quantization so far, for reasons related to renormalization. Another class of gauge theories with a non-Abelian gauge group, beginning with Yang–Mills theory, became amenable to quantization in the late 1960s and early 1970s, largely due to the work of Ludwig D. Faddeev, Victor Popov, Bryce DeWitt, and Gerardus 't Hooft. However, they remained very difficult to work with until the introduction of the BRST method. The BRST method provided the calculation techniques and renormalizability proofs needed to extract accurate results from both "unbroken" Yang–Mills theories and those in which the Higgs mechanism leads to spontaneous symmetry breaking. Representatives of these two types of Yang–Mills systems—quantum chromodynamics and electroweak theory—appear in the Standard Model of particle physics.
It has proven rather more difficult to prove the existence of non-Abelian quantum field theory in a rigorous sense than to obtain accurate predictions using semi-heuristic calculation schemes. This is because analyzing a quantum field theory requires two mathematically interlocked perspectives: a Lagrangian system based on the action functional, composed of fields with distinct values at each point in spacetime and local operators which act on them, and a Hamiltonian system in the Dirac picture, composed of states which characterize the entire system at a given time and field operators which act on them. What makes this so difficult in a gauge theory is that the objects of the theory are not really local fields on spacetime; they are right-invariant local fields on the principal gauge bundle, and different local sections through a portion of the gauge bundle, related by passive transformations, produce different Dirac pictures.
What is more, a description of the system as a whole in terms of a set of fields contains many redundant degrees of freedom; the distinct configurations of the theory are equivalence classes of field configurations, so that two descriptions which are related to one another by a gauge transformation are also really the same physical configuration. The "solutions" of a quantized gauge theory exist not in a straightforward space of fields with values at every point in spacetime but in a quotient space (or cohomology) whose elements are equivalence classes of field configurations. Hiding in the BRST formalism is a system for parameterizing the variations associated with all possible active gauge transformations and correctly accounting for their physical irrelevance during the conversion of a Lagrangian system to a Hamiltonian system.
The principle of gauge invariance is essential to constructing a workable quantum field theory. But it is generally not feasible to perform a perturbative calculation in a gauge theory without first "fixing the gauge"—adding terms to the Lagrangian density of the action principle which "break the gauge symmetry" to suppress these "unphysical" degrees of freedom. The idea of gauge fixing goes back to the Lorenz gauge approach to electromagnetism, which suppresses most of the excess degrees of freedom in the four-potential while retaining manifest Lorentz invariance. The Lorenz gauge is a great simplification relative to Maxwell's field-strength approach to classical electrodynamics, and illustrates why it is useful to deal with excess degrees of freedom in the representation of the objects in a theory at the Lagrangian stage, before passing over to Hamiltonian mechanics via the Legendre transformation.
The Hamiltonian density is related to the Lie derivative of the Lagrangian density with respect to a unit timelike horizontal vector field on the gauge bundle. In a quantum mechanical context it is conventionally rescaled by a factor
i\hbar
For perturbative purposes, we gather the configuration of all the fields of our theory on an entire three-dimensional horizontal spacelike cross section of P into one object (a Fock state), and then describe the "evolution" of this state over time using the interaction picture. The Fock space is spanned by the multi-particle eigenstates of the "unperturbed" or "non-interaction" portion
l{H}0
l{H}
l{H}0
Hence, in the zero-order approximation, the set of weights characterizing a Fock state does not change over time, but the corresponding field configuration does. In higher approximations, the weights also change; collider experiments in high-energy physics amount to measurements of the rate of change in these weights (or rather integrals of them over distributions representing uncertainty in the initial and final conditions of a scattering event). The Dyson series captures the effect of the discrepancy between
l{H}0
l{H}
To use the Dyson series to calculate anything, one needs more than a gauge-invariant Lagrangian density; one also needs the quantization and gauge fixing prescriptions that enter into the Feynman rules of the theory. The Dyson series produces infinite integrals of various kinds when applied to the Hamiltonian of a particular QFT. This is partly because all usable quantum field theories to date must be considered effective field theories, describing only interactions on a certain range of energy scales that we can experimentally probe and therefore vulnerable to ultraviolet divergences. These are tolerable as long as they can be handled via standard techniques of renormalization; they are not so tolerable when they result in an infinite series of infinite renormalizations or, worse, in an obviously unphysical prediction such as an uncancelled gauge anomaly. There is a deep relationship between renormalizability and gauge invariance, which is easily lost in the course of attempts to obtain tractable Feynman rules by fixing the gauge.
\partial\muA\mu=0
More sophisticated approaches do not attempt to apply a delta function constraint to the gauge transformation degrees of freedom. Instead of "fixing" the gauge to a particular "constraint surface" in configuration space, one can break the gauge freedom with an additional, non-gauge-invariant term added to the Lagrangian density. In order to reproduce the successes of gauge fixing, this term is chosen to be minimal for the choice of gauge that corresponds to the desired constraint and to depend quadratically on the deviation of the gauge from the constraint surface. By the stationary phase approximation on which the Feynman path integral is based, the dominant contribution to perturbative calculations will come from field configurations in the neighborhood of the constraint surface.
The perturbative expansion associated with this Lagrangian, using the method of functional quantization, is generally referred to as the Rξ gauge. It reduces in the case of an Abelian U(1) gauge to the same set of Feynman rules that one obtains in the method of canonical quantization. But there is an important difference: the broken gauge freedom appears in the functional integral as an additional factor in the overall normalization. This factor can only be pulled out of the perturbative expansion (and ignored) when the contribution to the Lagrangian of a perturbation along the gauge degrees of freedom is independent of the particular "physical" field configuration. This is the condition that fails to hold for non-Abelian gauge groups. If one ignores the problem and attempts to use the Feynman rules obtained from "naive" functional quantization, one finds that one's calculations contain unremovable anomalies.
The problem of perturbative calculations in QCD was solved by introducing additional fields known as Faddeev–Popov ghosts, whose contribution to the gauge-fixed Lagrangian offsets the anomaly introduced by the coupling of "physical" and "unphysical" perturbations of the non-Abelian gauge field. From the functional quantization perspective, the "unphysical" perturbations of the field configuration (the gauge transformations) form a subspace of the space of all (infinitesimal) perturbations; in the non-Abelian case, the embedding of this subspace in the larger space depends on the configuration around which the perturbation takes place. The ghost term in the Lagrangian represents the functional determinant of the Jacobian of this embedding, and the properties of the ghost field are dictated by the exponent desired on the determinant in order to correct the functional measure on the remaining "physical" perturbation axes.
Intuition for the BRST formalism is provided by describing it geometrically, in the setting of fiber bundles. This geometric setting contrasts with and illuminates the older traditional picture, that of algebra-valued fields on Minkowski space, provided in (earlier) quantum field theory texts.
In this setting, a gauge field can be understood in one of two different ways. In one, the gauge field is a local section of the fiber bundle. In the other, the gauge field is little more than the connection between adjacent fibers, defined on the entire length of the fiber. Corresponding to these two understandings, there are two ways to look at a gauge transformation. In the first case, a gauge transformation is just a change of local section. In general relativity, this is referred to as a passive transformation. In the second view, a gauge transformation is a change of coordinates along the entire fiber (arising from multiplication by a group element g) which induces a vertical diffeomorphism of the principal bundle.
This second viewpoint provides the geometric foundation for the BRST method. Unlike a passive transformation, it is well-defined globally on a principal bundle, with any structure group over an arbitrary manifold. That is, the BRST formalism can be developed to describe the quantization of any principle bundle on any manifold. For concreteness and relevance to conventional QFT, much of this article sticks to the case of a principal gauge bundle with compact fiber over 4-dimensional Minkowski space.
A principal gauge bundle P over a 4-manifold M is locally isomorphic to U × F, where U ⊂ R4 and the fiber F is isomorphic to a Lie group G, the gauge group of the field theory (this is an isomorphism of manifold structures, not of group structures; there is no special surface in P corresponding to 1 in G, so it is more proper to say that the fiber F is a G-torsor). The most basic property as a fiber bundle is the "projection to the base space" π : P → M, which defines the vertical directions on P (those lying within the fiber π−1(p) over each point p in M). As a gauge bundle it has a left action of G on P which respects the fiber structure, and as a principal bundle it also has a right action of G on P which also respects the fiber structure and commutes with the left action.
The left action of the structure group G on P corresponds to a change of coordinate system on an individual fiber. The (global) right action Rg : P → P for a fixed g in G corresponds to an actual automorphism of each fiber and hence to a map of P to itself. In order for P to qualify as a principal G-bundle, the global right action of each g in G must be an automorphism with respect to the manifold structure of P with a smooth dependence on g, that is, a diffeomorphism P × G → P.
ak{E}
Vak{E}
ak{E}
ak{g}
The "field theory" of interest is defined in terms of a set of "fields" (smooth maps into various vector spaces) defined on a principal gauge bundle P. Different fields carry different representations of the gauge group G, and perhaps of other symmetry groups of the manifold such as the Poincaré group. One may define the space
Pl
Pl0
\epsilon\inVak{E}
Identifying local gauge transformations with a particular subspace of vector fields on the manifold P provides a better framework for dealing with infinite-dimensional infinitesimals: differential geometry and the exterior calculus. The change in a scalar field under pullback along an infinitesimal automorphism is captured in the Lie derivative, and the notion of retaining only the term linear in the vector field is implemented by separating it into the interior derivative and the exterior derivative. In this context, "forms" and the exterior calculus refer exclusively to degrees of freedom which are dual to vector fields on the gauge bundle, not to degrees of freedom expressed in (Greek) tensor indices on the base manifold or (Roman) matrix indices on the gauge algebra.
The Lie derivative on a manifold is a globally well-defined operation in a way that the partial derivative is not. The proper generalization of Clairaut's theorem to the non-trivial manifold structure of P is given by the Lie bracket of vector fields and the nilpotence of the exterior derivative. This provides an essential tool for computation: the generalized Stokes theorem, which allows integration by parts and then elimination of the surface term, as long as the integrand drops off rapidly enough in directions where there is an open boundary. (This is not a trivial assumption, but can be dealt with by renormalization techniques such as dimensional regularization as long as the surface term can be made gauge invariant.)
Central to the BRST formalism is the BRST operator
sB
W(\deltaλ)
\deltaλ
sB
W(\deltaλ)X=\deltaλ sBX
X\in{Pl}0
{Pl}0
Like the exterior derivative, the BRST operator is nilpotent of degree 2, i. e.,
| 2 | |
| (s | |
| B) |
=0
sBX
\deltaλ
\iota\deltaλ.
\begin{align} \left[\iota\deltaλ,sB\right]sBX&=\iota\deltaλ(sBsBX)+sB\left(\iota\deltaλ(sBX)\right)\\ &=sB\left(\iota\deltaλ(sBX)\right) \end{align}
The Hamiltonian perturbative formalism is carried out not on the fiber bundle, but on a local section. In this formalism, adding a BRST-exact term to a gauge invariant Lagrangian density preserves the relation
sBX=0.
QB
[QB,l{H}]=0.
QB|\Psii\rangle=0
QB|\Psif\rangle ≠ 0
The nilpotence of the BRST operator can be understood as saying that its image (the space of BRST exact forms) lies entirely within its kernel (the space of BRST closed forms). The "true" Lagrangian, presumed to be invariant under local gauge transformations, is in the kernel of the BRST operator but not in its image. This implies that the universe of initial and final conditions can be limited to asymptotic "states" or field configurations at timelike infinity, where the interaction Lagrangian is "turned off". These states lie in the kernel of
QB,
QB,
QB
When defining the asymptotic states, the states that lie inside the image of
QB
The quantization prescription in the interaction picture is to build a vector space of BRST-closed configurations at a particular time, such that this can be converted into a Fock space of intermediate states suitable for Hamiltonian perturbation. As is conventional for second quantization, the Fock space is provided with ladder operators for the energy-momentum eigenconfigurations (particles) of each field, complete with appropriate (anti-)commutation rules, as well as a positive semi-definite inner product. The inner product is required to be singular exclusively along directions that correspond to BRST-exact eigenstates of the unperturbed Hamiltonian. This ensures that any pair of BRST-closed Fock states can be freely chosen out of the two equivalence classes of asymptotic field configurations corresponding to particular initial and final eigenstates of the (unbroken) free-field Hamiltonian.
The desired quantization prescriptions provide a quotient Fock space isomorphic to the BRST cohomology, in which each BRST-closed equivalence class of intermediate states (differing only by an exact state) is represented by exactly one state that contains no quanta of the BRST-exact fields. This is the appropriate Fock space for the asymptotic states of the theory. The singularity of the inner product along BRST-exact degrees of freedom ensures that the physical scattering matrix contains only physical fields. This is in contrast to the (naive, gauge-fixed) Lagrangian dynamics, in which unphysical particles are propagated to the asymptotic states. By working in the cohomology, each asymptotic state is guaranteed to have one (and only one) corresponding physical state (free of ghosts).
The operator
QB
To summarize: no field introduced as part of a BRST gauge fixing procedure will appear in asymptotic states of the gauge-fixed theory. However, this does not imply that these "unphysical" fields are absent in the intermediate states of a perturbative calculation! This is because perturbative calculations are done in the interaction picture. They implicitly involve initial and final states of the non-interaction Hamiltonian
l{H}0
sB
sB
T. Kugo and I. Ojima are commonly credited with the discovery of the principal QCD color confinement criterion. Their role in obtaining a correct version of the BRST formalism in the Lagrangian framework seems to be less widely appreciated. It is enlightening to inspect their variant of the BRST transformation, which emphasizes the hermitian properties of the newly introduced fields, before proceeding from an entirely geometrical angle.
The
ak{g}
G=\xi\partial\muA\mu,
\xi
ak{g}
A\mu.
ci
ak{g}
bi=\bar{c}i
ak{g}
Bi
ak{g}
The field
c
b
B
The BRST Lagrangian density is
l{L}=l{L}rm{matter}(\psi,A
| a) | |
| \mu |
-{1\over4g2}\operatorname{Tr}[F\mu\nuF\mu\nu]+{1\over2g2}\operatorname{Tr}[BB]-{1\overg2}\operatorname{Tr}[BG]-{\xi\overg2}\operatorname{Tr}[\partial\mubD\muc]
Here,
D\mu
A\mu.
c
Vak{E}
\deltaλ\inVak{E}
ak{g}
\psi
A\mu
c
While the Lagrangian density isn't BRST invariant, its integral over all of spacetime, the action is. The transformation of the fields under an infinitessimal gauge transformation
\deltaλ
\begin{align} \delta\psii&=\deltaλDic\\ \deltaA\mu&=\deltaλD\muc\\ \deltac&=\deltaλ\tfrac{i}{2}[c,c]\\ \deltab=\delta\bar{c}&=\deltaλB\\ \deltaB&=0 \end{align}
Note that
[ ⋅ , ⋅ ]
QB
\deltaλ
QB
QB=ci
| j} | |
| \left(L | |
| k |
bjck\right)
where
Li
fij{}k
\begin{align} QBA\mu&=D\muc\\ QBc&={i\over2}[c,c]\\ QBb&=B\\ QBB&=0 \end{align}
The details of the matter sector
\psi
\deltaA\mu
\partial\muA\mu=0
b=\bar{c}
B
\delta\bar{c}=i\deltaλB
\deltaλ
We already know, from the relation of the BRST operator to the exterior derivative and the Faddeev–Popov ghost to the Maurer–Cartan form, that the ghost
c
ak{g}
Vak{E}
-i(\partial\mu\bar{c})D\muc
\bar{c}
Vak{E}
ak{g}
\bar{c}
ak{g}
Vak{E}
B
ak{g}
\bar{c}
Vak{E}
A\mu.
B
ak{g}
Vak{E}
The one-particle states of the theory are discussed in the adiabatically decoupled limit g → 0. There are two kinds of quanta in the Fock space of the gauge-fixed Hamiltonian that lie entirely outside the kernel of the BRST operator: those of the Faddeev–Popov anti-ghost
\bar{c}
\bar{c}
sB
sB\left(\bar{c}\left(i\partial\muA\mu-\tfrac{1}{2}\xisB\bar{c}\right)\right).
Likewise, there are two kinds of quanta that will lie entirely in the image of the BRST operator: those of the Faddeev–Popov ghost
c
B
The anti-ghost is taken to be a Lorentz scalar for the sake of Poincaré invariance in
-i(\partial\mu\bar{c})D\muc
c
(Needs to be completed in the language of BRST cohomology, with reference to the Kugo–Ojima treatment of asymptotic Fock space.)
This section only applies to classical gauge theories. i.e. those that can be described with first class constraints. The more general formalism is described using the Batalin–Vilkovisky formalism.
The BRST construction [1] applies to a situation of a Hamiltonian action of a gauge group
G
M
{akg}
G
0\in{akg}*
\Phi:M\to{akg}*
| -1 | |
| M | |
| 0=\Phi |
(0)
G
M0
\tildeM
G
M0
The Hamiltonian mechanics of a gauge theory is described by
r
\Phii
M
M0
M0
M0
According to algebraic geometry, the set of smooth functions over a space is a ring. The Koszul-Tate complex (the first class constraints aren't regular in general) describes the algebra associated with the symplectic reduction in terms of the algebra
Cinfty(M)
First, using equations defining
M0
M
...\toK1(\Phi)\toCinfty(M)\to0
so that
H0(K(\Phi))=C
| infty(M | |
| 0) |
Hp(K(\Phi))=0
p\ne0
Then, for the fibration
M0\to\tildeM
| ⋅ | |
| (\Omega | |
| vert |
(M0),dvert)
| ⋅ | |
| \Omega | |
| vert |
(M0)
Λ ⋅ V* ⊗ Cinfty(\tildeM)
Λ ⋅ V*
V
Ki,j=ΛiV* ⊗ ΛjV ⊗ Cinfty(M).
Finally (and this is the most nontrivial step), a differential
sB
K= ⊕ i,jKi,j
dvert
K
| 2 | |
| (s | |
| B) |
=0
| 0 | |
| H | |
| sB |
=Cinfty(\tildeM)
with respect to the grading by the ghost number :
Kn= ⊕ i-j=nKi,j
Thus, the BRST operator or BRST differential
sB
There are two antiderivations,
\delta
d
sB
\delta+d+more
sB
s2=(\delta+d)2=\delta2+d2+(\deltad+d\delta)=0
Consider the supercommutative algebra generated by
Cinfty(M)
l{P}i
Cinfty(M)
\delta
\deltal{P}i=-\Phii
\deltaf=0
f\inCinfty(M)
| infty(M | |
| C | |
| 0) |
A longitudinal vector field on
M0
M0
p
p
d
\begin{align} d\omega(V0,\ldots,Vk)=&
| i | |
| \sum | |
| i(-1) |
d{
| Vi |
G
M
ak{g}
G
0\inak{g}*
ak{g})
\Phi:M\toak{g}*
| -1 | |
| M | |
| 0=\Phi |
(0)
G
M0
\widetildeM=M0/G
G
M0
\widetildeM=M//G
First, using the regular sequence of functions defining
M0
M
Λ\bullet{akg} ⊗ Cinfty(M).
The differential,
\delta
Cinfty(M)
Cinfty(M)
Λ\bullet{akg} ⊗ Cinfty(M)
{akg}\toCinfty(M)
S({akg})
Cinfty(M)
S(ak{g})
ak{g}
S({akg})\toCinfty(M)
ak{g}\toCinfty(M)
This Koszul complex is a resolution of the
S({akg})
Cinfty(M0)
Hj(Λ\bullet{akg} ⊗ Cinfty(M),\delta)=\begin{cases}Cinfty(M0)&j=0\ 0&j ≠ 0\end{cases}
Then, consider the Chevalley–Eilenberg complex for the Koszul complex
Λ\bullet{akg} ⊗ Cinfty(M)
ak{g}
K\bullet,\bullet=C\bullet\left(akg,Λ\bullet{akg} ⊗ Cinfty(M)\right)=Λ\bullet{akg}* ⊗ Λ\bullet{akg} ⊗ Cinfty(M).
The "horizontal" differential
d:Ki,\bullet\toKi+1,\bullet
Λ\bullet{akg} ⊗ Cinfty(M)
by the action of
ak{g}
Λ\bullet{akg}*
G
ak{g}
Let Tot(K) be a complex such that
\operatorname{Tot}(K)n=oplus\nolimitsi-j=nKi,j
with a differential D = d + δ. The cohomology groups of (Tot(K), D) are computed using a spectral sequence associated to the double complex
(K\bullet,\bullet,d,\delta)
The first term of the spectral sequence computes the cohomology of the "vertical" differential
\delta
| i,j | |
| E | |
| 1 |
=Hj(Ki,\bullet,\delta)=Λi{akg}* ⊗ Cinfty(M0)
The first term of the spectral sequence may be interpreted as the complex of vertical differential forms
| \bullet{\operatorname{vert}}(M | |
| (\Omega | |
| 0), |
d\operatorname{vert
M0\to\widetildeM
The second term of the spectral sequence computes the cohomology of the "horizontal" differential
d
| \bullet,\bullet | |
| E | |
| 1 |
| i,j | |
| E | |
| 2 |
\cong
| \bullet,j | |
| H | |
| 1 |
,d)=Cinfty
| g | |
| (M | |
| 0) |
=Cinfty(\widetildeM)
i=j=0
The spectral sequence collapses at the second term, so that
| i,j | |
| E | |
| infty |
=
| i,j | |
| E | |
| 2 |
Therefore,
Hp(\operatorname{Tot}(K),D)=Cinfty
| g | |
| (M | |
| 0) |
=Cinfty(\widetildeM)
Original BRST papers: