3.2 Supersymmetric theories on curved spaces from new minimal supergravity
4.1.3 Conformal Killing spinors are equivalent to conformal Killing vectors
In this section we study the geometrical constraints imposed by the existence of a charged conformal Killing spinor +. On the resulting curved backgrounds one can define a su-
perconformal field theory preserving some supersymmetry, see section 2.3. We will show that existence of a charged conformal Killing spinor is equivalent to the existence of a conformal Killing vector. In turn, this allows to introduce local coordinates, in which the metric (or equivalently the frame) takes a canonical form, generalising the one discussed in [102], corresponding to A = 0. We shall present this metric in section 4.2, in the case when the conformal Killing vector becomes a Killing vector. The two metrics are simply related by a Weyl rescaling.
First of all, notice that not only does a spinor + determine a null vector z (via (4.1)
or (4.2)), but also that in a sense the opposite is true. Indeed, let us study the map +7→ z. The space of spinors with fixed t++ is an S3 in the four-dimensional space of all
spinors. This is mapped by (4.2) into the space of all null vectors z with fixed z0, which is an S2 (the so-called “celestial sphere”). This is the Hopf fibration map, whose fibre is an S1. So, to any null vector z one can associate a U(1) worth of possible spinors + whose
bilinear is z.
Let us now move on to differential constraints. We consider the equation defining a charged conformal Killing spinor, or twistor-spinor,
∇A µ+ = 1 4γµD A + , (4.15)
where ∇Aµ = ∇µ − iAµ and DA is the covariantised Dirac operator DA = γµ∇Aµ. Note
that the equation does not mix chiralities, and we consider the case of a positive chiral spinor. A is a real connection and + is a section of the U(1) Hopf fibration described in
the previous paragraph.
We can expand equation (4.15) in the basis (4.9). Using (4.12), we obtain a set of linear equations for p, q and the gauge field A. Since the gamma-trace of equation (4.15) is trivial we find a total of six complex constraints
q · z = 0 , pA· e− = 0 , pA· z = 2q · ¯w ,
q · w = 0 , pA· ¯w = 0 , pA· w = −2q · e− , (4.16) where pA
µ = pµ− iAµ.
real conditions are constraints to be imposed on the geometry. We now show that these constraints are equivalent to the existence of a conformal Killing vector.
A short computation shows that
∇µzν = 2Re (pµzν+ 2 ¯qµwν) . (4.17)
Taking the anti-symmetric part of this equation we reproduce the first equation in (4.13). Taking the symmetric part and imposing that z is a conformal Killing vector,
(Lzg)µν = 2∇(µzν)= λgµν , (4.18)
we obtain the conditions
Re p · e− = 0 , q · z = q · w = 0 ,
Re p · z = 2Re(q · ¯w) , Re p · w = −q · e− , (4.19) with
λ ≡ 4 Re(q · ¯w) . (4.20)
This set of eight real conditions is precisely the subset of the constraints (4.16) not in- volving A, as previously stated.
To summarise, we showed that on any manifold M4 with a null conformal Killing
vector we can find a charged conformal Killing spinor. In section 4.1.5 we will give an expression for the gauge field A under which the conformal Killing spinor is charged.
Notice that the presence of a conformal Killing spinor also implies
dω = i2A − 3 ∗ (q ∧ e−∧ ¯w)∧ ω . (4.21) In the Euclidean signature case, we found a very similar condition (3.26), dω = W ∧ ω, where ω is the (2, 0) form of a complex structure. While in that case that condition turns out to be necessary and sufficient for the existence of a charged conformal Killing spinor, in the present case of Lorentzian signature this condition alone is not sufficient to imply supersymmetry.
4.1.4
Conformal Killing spinors are equivalent to conformal Killing–
Yano forms
Our two-form ω satisfies an interesting property, namely it is a charged conformal Killing– Yano form (CKF). In general, a p-form ϕ on a d-dimensional space(-time) (M, g) is con- formal Killing–Yano (or simply conformal Killing) if it satisfies the equation
∇ρϕµ1...µp = ∇[ρϕµ1...µp]+
p
d−p+1gρ[µ1∇
σ
This is a conformally invariant equation: if ϕ is a conformal Killing form on (M, g) and the metric is rescaled as g → ˜g = e2fg, then the rescaled form ˜ϕ = e(p+1)fϕ is conformal
Killing on (M, ˜g).
In the uncharged case (A = 0), it is known that the bilinears of conformal Killing spinors are conformal Killing forms, see e.g. [105]. We already saw that this is true for z, since a conformal Killing one-form is just the dual of a conformal Killing vector. For a two-form in four dimensions, this is easiest to check in the two-component formalism for spinors. Because of its definition in (4.1), ω can be written as ωαβ = αβ, and the CKF
equation (4.22) reads
Dα(β˙ ωγδ) = 0 . (4.23)
The CKS equation reads in this formalism D(βα˙ γ) = 0, which implies obviously (4.23). Since for us the CKS is actually charged under A, we obtain that (∇A)(βα˙ ωγδ) = 0, where ∇A
µ = ∇µ− 2iAµ; or, going back to four-component language,
∇A
ρωµν = ∇A[ρωµν]−23gρ[µ∇A σων]σ , (4.24)
which is a charged version of the standard conformal Killing form equation.
It is interesting to ask to what extent this property can be used to characterise our space-time, similarly to what we saw in section 4.1.3. First of all, we should ask when a two-form ω can be written as a spinor bilinear as in (4.1). One possible answer is that the form should define anR2 structure; namely, that the stabiliser of ω in SO(3,1) should
be R2. We can also give an alternative, more concrete characterisation by using again
the two-component formalism for spinors. The two-form ω should be imaginary self-dual, which means it is in the (1, 0) representation of SO(3,1); the corresponding bispinor then is a symmetric matrix ωαβ. As a 2 × 2 matrix, this can be factorised as αβ if and only
if it has rank 1, which is equivalent to det(ω) = 12αβγδωαγωβδ = 0; in the original form
language, ωµνωµν = 0. So we have obtained that a two-form ω can be written as a bispinor
as in (4.1) if and only if it is imaginary self-dual and null:
ω = +⊗ − ⇐⇒ ( ∗ω = iω , ωµνω µν = 0 . (4.25)
Remarkably, it turns out that the content of the equation (4.24) for a CKF is exactly the same as the content of the system (4.15) for a CKS. Indeed, if one uses (4.12) in (4.24) (or in (4.23)), the system one finds is exactly (4.16). Thus, we can conclude that a choice of metric and gauge field A admits a charged CKS if and only if it admits a null, imaginary self-dual charged CKF.
This reformulation is slightly less interesting than the one involving a null CKV in section 4.1.3. Although CKF’s do have physical applications (such as helping in finding first integrals of the geodesic equation, see for example [106]), their geometrical meaning is less compelling than that of a CKV. Moreover, one needs both the data of the geometry and of the gauge field A to check the condition (4.24), whereas in the previous section we saw that the presence of a null CKV tells us that a geometry can admit a charged CKS for some A (without having to guess its form, which will actually be determined in section 4.1.5). Last but not least, the CKV condition is computationally easier to check than the CKF condition (4.24).