apart from the ones suppressed by Λ5. The resulting Lagrangian takes the form Ldec = Lkin(˜hµν, Aµ, χ)− 1
18Λ55
˜hµν∂µ∂α∂βχ∂ν∂α∂βχ−˜hµν∂µ∂α∂νχ∂αχ
−1
2h∂˜ µ∂α∂βχ∂µ∂α∂βχ+ 1
2˜h∂αχ∂αχ
+ 1
432Λ55
(2 + 8k1+ 16k2+ 32k3)(χ)3+ (2−24k1−16k2)χχ2χ +(1−12k1−8k2)χ23χ
, (6.44)
whereLkin contains the kinetic terms of all helicities (6.40) with mass equal to zero due to the decoupling limit.
On the equations of motion, one can then use the freedom in the parametersk1, k2, k3to eliminate the higher derivative self-interactions of χ. One can already infer from (6.44) that it is not possible to eliminate all higher derivative interactions. In order to cancel theχ self-interactions we choose
1 + 4k1+ 8k2+ 16k3 = 0
1−12k1−8k2 = 0. (6.45)
The remaining interactions at the scale Λ5 L(2+3)dec5 =Lkin(˜hµν, Aµ, χ)− 1
36Λ55h˜ρσEρσµν∂µ∂αχ∂ν∂αχ (6.46) indicate the appearance of an additional degree of freedom on a background for ˜hµν
[76]. However, for perturbative backgrounds which are accessible within the effective description, the ghost mass always remains above the cutoff. Hence, the ghost does not appear in the physical spectrum. The stipulated UV completion must then take care of this possible instability. Within the framework of the gravitational Higgs mechanism, a similar instability for the Lagrangian proposed in [72] was found at fourth order in [267].
Without sources, the above conclusion, however, only remains true if the theory is truncated at the cubic level. As has later been studied in [74], adding the fourth order interaction in hµν shows that the interactions at Λ5 are given by a redundant operator which thus can be removed by a field redefinition. The fourth order interactions are actually the only other interactions which come with the scale Λ5 and are given by
1
362Λ105 ∂µ∂ρφ∂ρ∂νφEµναβ∂α∂σφ∂β∂σφ . (6.47) Now, one can perform a nonlinear field redefinition of the field ˜hµν of the following form
˜hµν = ¯hµν+ 1
36Λ55∂µ∂αχ∂ν∂αχ (6.48)
which yields a free cubic Lagrangian for the scale Λ5. Note that this field redefinition is canonical and invertible. The Lagrangian after the field redfinition is
L= 1 2
h¯µνEµναβ¯hαβ −1
8FµνFµν+ 1
12χχ . (6.49)
Thus, we recover a free theory at the scale Λ5. As pointed out in [74], all scales lower than Λ3 are redundant operators if one considers all higher order interaction terms appearing with the respective scale. In terms of helicities, they can be removed by appropriately defined field redefinitions which are always local and invertible suggesting that they cannot change the number of degrees of freedom.
Let us make a very important observation here. We have seen that all scales below Λ3
are in fact redundant. Therefore, the true strong coupling scale of the theory is Λ3. However, it is not possible to find an expansion in terms of Λ3 of the operators hµν which can be truncated at a given order. To be more precise, let us not consider the theory in terms of helicities or St¨uckelberg fields but rather in terms of the spin-2 field hµν corresponding to what was called unitary gauge in section 6.5.1. For instance, the theory considered as an expansion in powers ofhµν can only then have Λ5as a redundant coupling if all appropriate powers ofhµν are included. Even though the theory is weakly coupled at Λ5when including both cubic and quartic interactions ofhµν, it is not possible to truncate the theory at cubic order in hµν, because this would reintroduce the strong coupling scale Λ5. The reason for this is that the helicity-0 polarizations, which are the part ofhµν ∝ kmµk2ν, are already strongly coupled at the scale Λ5 and therefore one needs to consider all higher order interactions suppressed by this scale. If the coefficients are tuned appropriately, they combine to redundant operators.
6.7.2 Coupling to Sources
Let us now analyse how the coupling to sources can affect the above conclusions. In order to be as concrete as possible, first, consider the following example Lagrangian, also discussed in [74]
L= 1
2φφ+1
2ψψ+ψ(∂µ∂νφ)2
Λ5 +(∂µ∂νφ)2(∂α∂βφ)2
Λ10 . (6.50)
In [74] it was demonstrated that though seemingly a higher derivative action, four initial conditions suffice to solve the corresponding equations of motion, indicating the absence of additional degrees of freedom.
Performing a Hamiltonian analysis of (6.50) leads to the same conclusion. Introducing two auxiliary fields µ = ¨φ and ρ = ˙φ to account for higher derivatives allows for a straightforward counting of constraints. While initially the phase space dimension is enlarged from four to eight, one first class and two second class constraints are found which removes four phase space dimensions, cf. [74]. Thus, only two of the four fields are propagating.
At the same time, however, introducing sources into the action can lead to very different conclusions. A linear couplingJφφ+Jψψreintroduces the ghost problem into the action as was also discussed in [74]. In the Hamiltonian analysis, this is reflected by the fact that a linear coupling to sources converts the first class constraint into a second class constraint, while at the same time eliminating the tertiary second class constraint, see [74]. A system of two second class constraints can only reduce the phase space dimension by two in our case from eight to six, indicating the presence of the ghost. On the other hand, specifying a coupling of the form
Jφφ+Jψ
ψ+ 1
Λ5(∂µ∂νφ)2
(6.51) avoids this problem. This can again be confirmed in a Hamiltonian analysis of the full action, where both the first class and the tertiary constraint are now preserved.
For a classical analysis, cf. section 6.4, to detect the possible appearance of ghost-like degrees of freedom on certain backgrounds, it is sufficient to prove the existence of very mild “-backgrounds” that induce additional poles of the propagator. It is straightforward to see that a weak background of the formψ=results in the following kinetic action forφ:
1 2φ
φ+ 2 Λ52φ
. (6.52)
While the resulting ghost obviously has a mass which is parametrically larger than the cut-off Λ, this is of no immediate interest. Eq. (6.52) simply proves the existence of backgrounds with additional degrees of freedom.
How can this be reconciled with the previously cited Hamiltonian analysis for (6.50)?
The answer is quite simple. Specifying the coupling to sources makes implicit statements about the physical degrees of freedom of the system. With a linear coupling, the fields φ and ψ are the physical degrees of freedom. They can be excited independently. A background ψ = is a valid physical configuration and leads to a higher derivative action for φ. The additional pole seen in (6.52) corresponds to the additional modes found in a Hamiltonian analysis of (6.50) with a linear coupling to sources. On the other hand, specifying a coupling of the form (6.51) implies thatφand ψ+Λ15(∂µ∂νφ)2 are the physical degrees of freedom. This immediately signals that a nonlinear field redefinition should be performed to simplify an analysis of the properties of the theory.
Furthermore, the nonlinear coupling to sources has to be explicitly taken into account for the ghost analysis. Choosing a background ψ=, and accordingly Jψ =−, leads to a contribution to the quadratic Lagrangian from the source coupling that exactly cancels the higher derivative term in (6.52).
We have now understood that a ghost analysis relies crucially on specifying the physical degrees of freedom and this is exactly where shortcomings of a Hamiltonian analysis be-come visible. Performing such an analysis without specifying the coupling to sources is not sufficient to exclude the appearance of ghosts on viable backgrounds. The Hamilto-nian analysis of (6.50) allowed one to conclude that the system is free of ghosts, seemingly contradicting a straightforward stability analysis. However, we have seen that the latter
includes additional information, which, when correctly translated into the Hamiltonian, leads to an agreement between both methods.
We shall now apply this reasoning to the scenario of massive Einstein gravity. While it is immediately clear that a Hamiltonian analysis without specifying a coupling to sources does not capture the correct physics on arbitrary backgrounds, the question whether one can restrict couplings to sources, as addressed in [74], is much more subtle.
Restrictions in this case consist of explicit conditions on the sources, such as demanding covariant conservation. It is thus not as straightforward to identify the physical degrees of freedom as in the previous case, where only the linearly coupled fields deserved the name of propagating degrees of freedom.
In order for the coupling Λ5 in [72, 74] to remain redundant when introducing sources Tµν, it is important also to include the nonlinear couplings to the sources. Assuming the source is covariantly conserved, √
−g∇µTµν ≡ ∂µ(√
−gTµν) +√
−gΓνµγTµγ = 0, where ∇µ is the covariant derivative with respect to the full metric gµν and Γνµγ =
1
2gνδ(∂µgδγ+∂γgδµ−∂δgµγ) are the Christoffel symbols of the metric. The fact that the sources are only covariantly conserved is a consequence of the backreaction of gravity on matter, which is is equivalent to considering an infinite power series of nonlinear couplings hnµνTµν, where n ∈ N. Once again, it is then important that one considers the higher order terms corresponding to the covariant conservation if Λ5 is to be a redundant coupling. It has been shown in [74] that if one does so, one can indeed retain this redundancy. For this reason, in terms of the helicities, the field redefinition (6.48) does not introduce any higher derivatives in the coupling to sources.
Let us briefly note that on the classical level one can always choose sources such that the condition of covariant conservation is fulfilled. In contrast to gravity however in massive gravity there is no symmetry a priori which would protect the covariant conservation of sources. Therefore, for a full quantum analysis one has to make sure that radiative corrections leave the coupling Λ5 redundant. In [74] it was suggested that this is indeed the case.
We want to end this section with a general comment on nonlinear field redefinitions.
While well-defined, invertible field redefinitions obviously just correspond to a renaming of variables and cannot change the physical content of a self-contained theory, these properties have to be carefully checked. Furthermore, once the observable degrees of freedom of a theory are specified, a coupling to sources has to be taken into account, as we have seen in the example given above. Otherwise, nonlinear redefinitions, even if invertible, may change the notion of physical degrees of freedom and can thus give misleading results.