symmetry
S S
Article
Space, Matter and Interactions in a Quantum Early Universe
Part I: Kac–Moody and Borcherds Algebras
Piero Truini 1,2, Alessio Marrani 3,*
, Michael Rios 2 and Klee Irwin 2






Citation: Truini, P.; Marrani, A.; Rios,
M.; Irwin, K. Space, Matter and
Interactions in a Quantum Early
Universe Part I: Kac–Moody and
Borcherds Algebras. Symmetry 2021,
13, 2342. https://doi.org/10.3390/
sym13122342
Academic Editors: Lucrezia Ravera
and Vasilis K. Oikonomou
Received: 29 October 2021
Accepted: 1 December 2021
Published: 6 December 2021
Publisher’s Note: MDPI stays neutral
with regard to jurisdictional claims in
published maps and institutional affil-
iations.
Copyright: © 2021 by the authors.
Licensee MDPI, Basel, Switzerland.
This article is an open access article
distributed under the terms and
conditions of the Creative Commons
Attribution (CC BY) license (https://
creativecommons.org/licenses/by/
4.0/).
1
Istituto Nazionale Fisica Nucleare (INFN) , Sez. di Genova, Via Dodecaneso 33, I-16146 Genova, Italy;
piero.truini@ge.infn.it
2 Quantum Gravity Research (QGR), 101 S. Topanga Canyon Rd., 1159, Los Angeles, CA 90290, USA;
mrios@dyonicatech.com (M.R.); Klee@quantumgravityresearch.org (K.I.)
3 Centro Studi e Ricerche Enrico Fermi, Via Panisperna 89A, I-00184 Roma, Italy
* Correspondence: jazzphyzz@gmail.com
Abstract: We introduce a quantum model for the universe at its early stages, formulating a mechanism
for the expansion of space and matter from a quantum initial condition, with particle interactions and
creation driven by algebraic extensions of the Kac–Moody Lie algebra e9. We investigate Kac–Moody
and Borcherds algebras, and we propose a generalization that meets further requirements that we
regard as fundamental in quantum gravity.
Keywords: Kac–Moody algebras; Borcherds algebras; early universe cosmology
1. Introduction
This is the first of two papers—see also [1]—describing an algebraic model of quan-
tum gravity.
The intrinsic difficulty of quantizing gravity, encountered also in the most acclaimed
approaches of string theory and loop quantum gravity, has led us to this attempt of thinking
outside the box and exploiting only the most fundamental principles of quantum mechanics
and general relativity, as we believe that they should apply in the extreme conditions of a
hot dense universe in its early stages. We use, therefore, an Occam’s razor type of approach
by starting with the least possible assumptions and filling in more structure when needed.
Two fundamental and intuitive physical principles are assumed to hold in the regime
we study:
FP1 There is no classical observable to be quantized. We start directly from quantum
objects and states of the system: interactions, creation and the expansion of spacetime
all occur with quantum probability amplitudes; the universe evolves from an initial
quantum state. Gravity is identified with the way spacetime is created as particles
evolve, whereas the rest of the dynamics, which incorporate electro-weak and strong
forces, only depend on the quantum charges of the constituents and the rule for
the building blocks of the interactions, which we assume to be algebraic in nature.
We depart from the conventional view that quantum gravity should be realized as
the quantization of gravity with its renormalization, with the four fundamental forces
unifying at the Planck scale. We are at the Plank scale: the charges of the constituents
are symmetrical, but diversity comes from quantum theory itself, and from the
initial conditions.
FP2 There is no spacetime geometry to start with. We can only start by establishing one
rule for the interactions and one for the creations of spacetime, bearing in mind that
they both occur with probability amplitudes: both interactions and spacetime have
an intrinsic quantum nature.
The assumption, supported by physical observations, that at very high temperatures
the interactions are tree-like, has led us to consider algebraic models, with the algebra
Symmetry 2021, 13, 2342. https://doi.org/10.3390/sym13122342
https://www.mdpi.com/journal/symmetry
Symmetry 2021, 13, 2342
2 of 25
product playing the role of the building blocks of all interactions. A mechanism
for the quantum creation of spacetime suggests the inclusion of momenta within
the charges (roots) of the algebra, thus achieving both charge and energy–momentum
conservation. We reach this goal by considering infinite dimensional generalized Lie
algebras. A generator in the algebra is related to a particle, with certain charges
coming from the algebra roots, but it is also related to a quantum field, since new
generators are produced by multiplication in the algebra: a generator expands in
spacetime with complex quantum amplitudes, but locally interacts, disappears and
contributes to the creation of new generators. This local action can be considered
as a vertex, made of generators obeying the rules of an algebra. There is no vacuum,
since space points exist only where generators are. We see, therefore, that we do have
an algebra at the core of the model, but the expansion of spacetime embeds it into a
larger picture: that of a vertex-type algebra describing quantum interactions and a
quantum-generated spacetime.
We believe that the above considerations are plausible and strongly based on fun-
damental physics. The evolution of the universe, its quantum interactions and quantum
expansion from a chosen initial state of a finite set of generators, can be turned into an
algorithm in which all physical quantities are, in principle, calculable, and no infinity
occurs. A concrete, calculable realization of the above ideas is what we have achieved
with our model, which has no claim other than being physically consistent and mathemati-
cally rigorous.
In this first paper, we provide an introduction to the foundations of the model, and we
start investigating the mathematical structures that suit our purpose. In the second paper [1],
we will deal with a physical model relying on a particular infinite dimensional algebra.
The Lie algebra at the core of our model has the following features and interpretation:
1.
It is an infinite dimensional Lie algebra extending e8 that is regarded as the internal
quantum number subalgebra, meaning that the e8 roots represent the charges and
spin of elementary particles;
2.
Its root lattice is Lorentzian;
3.
The subspace of the lattice that is complementary to that of e8 is interpreted as
momentum space.
Remark 1. The Lie algebra e8 has been considered by many as a possible algebra for grand
unification, as well as for quantum gravity. It has then been considered not suitable after the no-go
theorem by Distler and Garibaldi [2]. We will show in Section 2.3 how e8 fulfils the requirements for
standard model degrees of freedom and algebras, which seems to contradict the thesis of Distler and
Garibaldi. We underline here that it does not, since the hypothesis, denoted TOE1 by the authors
of [2]—in particular, the fact that the algebra of the standard model centralizes sl(2,C)—not
only does not apply, but actually needs not to do so, as will become obvious in the development of
Section 2.3.
Algebraic methods are extensively used and successfully exploited in string theory
and conformal field theory in two dimensions, through the concept of vertex operator
algebras, [3–5], in order to describe the interactions between different strings, localized at
vertices, analogously to the Feynman diagram vertices. Mathematically, the underlying
concept of a vertex algebra was introduced by Borcherds [6–8], in order to prove the
Monstrous Moonshine conjecture [9].
Infinite dimensional Kac–Moody algebras have recently entered the loop quantum
gravity literature to describe spin network edge modes (generalizing the Gibbons–Hawking
boundary term); e.g., [10,11], in which a boundary SU(2) symmetry was found, along with
a Virasoro structure that resembles strings with an internal three-dimensional structure.
On the other hand, the Kac–Moody algebra e11 has been investigated in string theory and
M-theory by P. West and collaborators (cf. e.g., [12] and Refs. therein). As we will elucidate
further below, we consider a Borcherds extension of the even larger Kac–Moody algebra
Symmetry 2021, 13, 2342
3 of 25
e12 in order to describe all interactions in a very early universe, and texploit the Grassmann
envelop in order to deal with fermions and implement Pauli’s exclusion principle.
The algebras used in this paper may be regarded as vertex operator algebras in a broader
sense, since they are characterized by interaction operators that look like generators of a Lie
algebra, and whose product depends upon parameters related to the spacetime creation
and expansion. The Lie algebra acts locally, but it is immersed in a wider, vertex-type
algebra by means of a mechanism that creates a discrete quantum spacetime.
The Pauli exclusion principle is fulfilled by turning the algebra into a Lie superalgebra
using the Grassmann envelope.
The resulting model is thus intrinsically relativistic, both because of the way spacetime
expands and because the Poincaré group acts locally on the Lie algebra. Furthermore,
the conservation of charge and momentum is a consequence of the Lie product, and, in this
respect, they are treated at the same level.
1.1. gu, a Lie Algebra for Quantum Gravity
At a very fundamental level, we make the following assumptions on quantum gravity,
founded on the current theoretical and experimental knowledge in physics.
(QG.1)
Gravity is a characteristic of spacetime;
(QG.2)
Spacetime is dynamical and related to matter. Therefore, we assume that it
emerges from the existence of particles and their interactions. There is no way of
defining distances and time lapses without interactions, so that the creation and
expansion of spacetime is itself a rule followed by particle interactions;
(QG.3)
A suitable mathematical structure at the core of the description of quantum
gravity is that of an algebra, which we will henceforth denote by gu, whose
generators represent the particles and whose product yields the building blocks
of the interactions (let us call them elementary interactions). As a consequence,
the interactions are endowed with a tree structure, thus opening up the opportu-
nity for a description of scattering amplitudes in terms of what we would call
gravitahedra, providing a generalization of the associahedra and permutahedra
in the current theory of scattering [13–20]. The structure constants of the algebra
determine the quantum amplitudes of the elementary interactions; in particular,
we assume gu to be a Lie algebra because it enables us to derive the fundamental
conservation laws observed in physics directly from the action of the genera-
tors as derivations (Jacobi identity). As in the theory of fields, the interactions
may only occur locally, point-by-point in the expanding spacetime, which can
therefore be viewed as a parameter on which the algebra product depends;
(QG.4)
In agreement with the theory of a big bang, strongly supported by the current ob-
servations, we assume the existence of an initial quantum state, mathematically
represented by an element of the universal enveloping algebra of gu. Such an ele-
ment is made of generators that can all interact among themselves, thus yielding
the first geometrical interpretation: that of a point where particles may interact;
(QG.5)
A particle has a certain probability amplitude to interact but also not to interact,
in which case, it expands, as described in Section 1.2;
(QG.6)
Particles are quantum objects, hence their existence through interactions occurs
with certain amplitudes. Therefore, spacetime acquires a quantum structure: a
point in space and time where particles are present with a certain amplitude and
may interact. The amplitude related to the quantum spacetime point is the sum
of the amplitudes for particles to be there. Consequently, the fact that gravitation
appears as an attractive force has to be explained through amplitudes and their
interference;
(QG.7)
The initial set of generators is finite by assumption, being the the initial state
represented by an element of the universal enveloping algebra. These gener-
ators are all allowed to interact with each other with a certain amplitude and
according to the algebra relations, at what we call time 0 of the universal clock.
Symmetry 2021, 13, 2342
4 of 25
The outcome of the first finite number of interactions, plus the creation of space,
which is a consequence of the momentum part of the root associated with each
generator, leads to a second finite set of interactions, and so on. What we call
universal time is this order parameter of the interactions. The expansions are also
countable, hence discrete: the structure of spacetime that emerges is discrete
and finite at every instant of the universal clock, as is the universe and the
quantum theory describing it. There is no divergence of any sort: quantum field
theory in the continuum, with its divergences and related renormalizations, is
an approximation that may be useful for calculations long after the big bang;
(QG.8)
The finiteness of the expanding universe, and thus the absence of spacetime
beyond it, affects the quantum initial state of particles, which are not free to move
on the spacetime stage but are bound as if they were surrounded by infinitely
high barriers. The steady state of such a particle is a superposition of states
with opposite 3-momenta, representing an object that moves simultaneously in
opposite directions, where, by 3-momentum, we denote the spatial component
of 4-momentum.
1.2. Expansion
The assignment of opposite 3-momenta is inherent to the quantum behavior of a
particle in a box, in which, the square of the momentum, but not the momentum itself,
has a definite value in a stationary state. In standard relativistic and non-relativistic
quantum mechanics, the ground state is a superposition of generalized states, with opposite
momenta ei (kx+a) and e−i (kx+a). We maintain the same energy and start enlarging the
box on opposite sides along the direction of ~p in steps of |~p|/E in Planck units, so that
a massless particle travels at velocity c, and a massive one travels slower than that. We
obtain a wave proportional to sin(
π
2
(x/a + n)) for a = 1 = |~p|/E and n = 1, 2, 3, . . ..
The wave function for the first four expansions is shown in Figure 1. We take the
discretized picture of the sine function maxima and minima, the dots, with cosmological
time t = n− 1.
~p/E
amp.
t = 0
t = 1
t = 2
t = 3
Figure 1. Expansion of a particle (blue dot) along ~p (x-axis) with amplitudes (y-axis).
The amplitude also acquires a time-dependent phase ei Et that makes it complex.
1.3. Fermions and Bosons
The Lie algebras considered in this paper contain e8, and thus d8. Under the adjoint
action of d8, the generators of e8 split into spinorial and non-spinorial ones, providing
the algebra with a two-graded structure. We give the spinorial generators the physical
Symmetry 2021, 13, 2342
5 of 25
meaning of fermions, whereas the non-spinorial generators are given the physical meaning
of bosons, in order to automatically comply with the addition of angular momenta.
On the other hand, the Pauli exclusion principle is embodied in the Grassmann envelope,
which turns the two-graded algebra into a Lie superalgebra. The degrees of freedom of the
spin-1/2 fermions originate from the superposition of opposite 3-momenta and the corre-
sponding change in helicity caused by the reflection at the space boundary. The Poincaré
group then naturally emerges as a group of transformations of the local algebra gu, leaving
the charges invariant.
All of these topics will be treated in the companion paper [1].
1.4. Quantum Quasicrystal
The expansion of the space that we propose has two fundamental features:
1.
A space point may exist with a certain probability amplitude, this latter being the sum
of the amplitudes for some particles—matter or radiation—to be there: no space point
can possibly be empty;
2.
Space is a quantum object that expands according to algebraic rules.
As a result of these two features, our model of the universe can be conceived as a
quantum quasicrystal [21–23].
2. e8, the Charge/Spin Subalgebra
In our treatment, we use the following labels for the Dynkin diagram of e8:
e
e
e
e
e
e
e
e
α1
α2
α3
α4
α5
α7
α8
α6
(1)
A way of writing the simple roots of e8 in the orthonormal basis {k1, . . . , k8} of R8 is:
α1 =
k1 − k2;
α2 =
k2 − k3;
α3 =
k3 − k4;
α4 =
k4 − k5;
(2)
α5 =
k5 − k6;
α6 =
k6 − k7;
α7 =
k6 + k7;
α8 = −
1
2
(k1 + k2 + k3 + k4 + k5 + k6 + k7 + k8).
The whole root system Φ8 of e8 (obtained from the simple roots by Weyl reflections) can be
written as follows:
Φ8 = ΦO ∪ΦS (240 roots);
ΦO :=
{
±ki ± k j|1 6 i < j 6 8
}
,
4(82) = 112;
ΦS :=
{
1
2 (±k1 ± k2 ± ...± k7 ± k8), even # of +
}
,
27 = 128.
(3)
The first set ΦO of 112 roots is the set of roots of d8 ' so(16,C). The set ΦS is a Weyl spinor
of d8, with respect to the adjoint action (every orthogonal Lie algebra in even dimension
dn ' so(2n,C) has a Weyl spinor representation of dimension 2n−1).
If α is a root, there is a unique way of writing it as α = ∑ λiαi, where the αi’s are simple
(in fact, all λi’s are positive for positive roots, and negative for negative roots). The sum
ht(α) := ∑ λi is called the height of α.
Symmetry 2021, 13, 2342
6 of 25
The fact that the roots of ΦO are the roots of a subalgebra and those of ΦS correspond
to a representation of it can be seen by noticing that: ΦO +ΦO ⊂ ΦO , ΦO +ΦS ⊂ ΦS .
Moreover, ΦS +ΦS ⊂ ΦO implies that ΦO is embedded into e8 in a symmetric way.
Thus, one can consistently define a non-Cartan generators xα bosonic if α ∈ ΦO , and
fermionic if α ∈ ΦS . We also call fermionic or bosonic the root α associated to a fermionic
(resp. bosonic) non-Cartan generator xα. A Cartan generator hα is always bosonic for any
α, since hα = [x−α, xα]. The roots of e8 split into 128 fermions (F) and 112 bosons (B).
2.1. Algebraic Structure
The e8 algebra can be defined from its root system Φ8 [24–26], over the complex field
extension C of the rational integers Z in the following way:
(a)
We select the set of simple roots ∆8 of Φ8;
(b)
We select a basis {h1, ..., h8} of the eight-dimensional vector space h over C and set
hα = ∑8
i=1 λihi for each α ∈ Φ8, such that α = ∑
8
i=1 λiαi;
(c)
We associate to each α ∈ Φ8 a one-dimensional vector space Lα over C spanned
by xα;
(d)
We define e8 = h
⊕
α∈Φ8
Lα as a vector space over C;
(e)
We give e8 an algebraic structure by defining the following multiplication on the ba-
sis {h1, . . . , h8}∪ {xα | α ∈ Φ8} by linearity to a bilinear multiplication e8 × e8 → e8:
[hi, hj] = 0 , 1 ≤ i, j ≤ 8
[hi, xα] = −[xα, hi] = (α, αi) xα , 1 ≤ i ≤ 8 , α ∈ Φ8
[xα, x−α] = −hα
[xα, xβ] = 0 for α, β ∈ Φ8 such that α+ β /∈ Φ8 and α
6= −β
[xα, xβ] = ε(α, β) xα+β for α, β ∈ Φ8 such that α+ β ∈ Φ8
(4)
where ε(α, β) is the asymmetry function, introduced in [27], as in Definition 1; see
also [26].
Definition 1. Let Le8 denote the lattice of all linear combinations of the simple roots with inte-
ger coefficients
Le8 =
{
8
∑
i=1
ciαi | ci ∈ Z , αi ∈ ∆8
}
(5)
the asymmetry function ε(α, β) : Le8 ×Le8 → {−1, 1} is defined by:
ε(α, β) =
8
∏
i,j=1
ε(αi, αj)
`imj
for α =
8
∑
i=1
`iαi , β =
8
∑
j=1
mjαj
(6)
where αi, αj ∈ ∆8 and
ε(αi, αj) =

−1
if i = j
−1
if αi + αj is a root and αi < αj
+1 otherwise
(7)
We recall the following standard result on the roots of e8 (normalized to 2), [24,25]:
Proposition 1. For each α, β ∈ Φ8, the scalar product (α, β) ∈ {±2,±1, 0}; α+ β (respectively,
α− β) is a root if and only if (α, β) = −1 (respectively, +1); if both α+ β and α− β are not in
Φ8 ∪ {0}, then (α, β) = 0.
For α, β ∈ Φ8, if α+ β is a root, then α− β is not a root.
Symmetry 2021, 13, 2342
7 of 25
The following properties of the asymmetry function follow from its definition [26].
Proposition 2. The asymmetry function ε satisfies, for α, β,γ ∈ Le8 :
(i)
ε(α+ β,γ) =
ε(α,γ)ε(β,γ)
(ii)
ε(α, β+ γ) =
ε(α, β)ε(α,γ)
(iii)
ε(α, α) = (−1) 12 (α,α) ⇒ ε(α, α) = −1 if α ∈ Φ8
(iv)
ε(α, β)ε(β, α) = (−1)(α,β) ⇒ ε(α, β) = −ε(β, α) if α, β, α+ β ∈ Φ8
(v)
ε(0, β) =
ε(α, 0) = 1
(vi)
ε(−α, β) =
ε(α, β)−1 = ε(α, β)
(vii)
ε(α,−β) =
ε(α, β)−1 = ε(α, β)
Property (iv) shows that the product in (4) is indeed antisymmetric.
2.2. e8 Charges and the Magic Star
There are four orthogonal a2’s in e8, where orthogonal means that the planes on which
their root systems lie are orthogonal to each other.
We denote one of them ac2 for color, one a
f
2 for flavor and the other two as a
(1)
2 and a
(2)
2
:
ac2 : ki − k j , i
6= j , i, j ∈ {1, 2, 3}
a
f
2 : ki − k j , i
6= j , i, j ∈ {4, 5, 6}
a
(1)
2
: ±(k7 + k8), ± 12 (k1 + k2 + k3 + k4 + k5 + k6 − k7 − k8)
± 12 (k1 + k2 + k3 + k4 + k5 + k6 + k7 + k8)
a
(2)
2
: ±(k7 − k8) , ± 12 (−k1 + k2 − k3 + k4 − k5 + k6 − k7 + k8)
± 12 (−k1 + k2 − k3 + k4 − k5 + k6 + k7 − k8)
The generators of ac2 and a
f
2 are bosonic.
The Magic Star (MS) of e8 shown in Figure 2 is obtained by projecting its roots
on the plane of ac2 [28]. The pair of integers {r, s} are the (Euclidean) scalar products
r := (α, k1 − k2) and s := (α, k1 + k2 − 2k3) for each root α. The fermions on the tips of the
MS are quarks, since they are acted upon by ac2: they are colored. The fermions within the
center of the MS are leptons: they are colorless. A similar MS of e6 within e8 is obtained by
projecting the roots in the center of the MS of e8 on the plane of a
f
2 .
go
{2,0}
{-2,0}
{1,3}
{-1,3}
{-1,-3}
{1,-3}
{1,1}
{-1,1}
{-1,-1}
{1,-1}
{0,2}
{0,-2}
Figure 2. The Magic Star (MS): g0 = e6 in the MS of e8, g0 = a
(1)
2 ⊕ a
(2)
2
in the MS of e6; the triangles
represent the 3 and 3 representations of the a2 with roots in the external hexagon.
Symmetry 2021, 13, 2342
8 of 25
Notice that, in each tip of the MS of e8, we obtain 27 roots, 11 of which are bosonic
and 16 fermionic; this corresponds to the following decomposition of the irrepr. 27 of e6
e6 ⊃ d5 ⊕C,
27 = 1−4 ⊕ 102 ⊕ 16−1.
(8)
On the other hand, within e6, we have nine roots in each tip of the MS, five of which are
bosonic and four fermionic; this corresponds to the following decomposition of the repr.
(3, 3) of a2 ⊕ a2 :
a2 ⊕ a2 ⊃ a1 ⊕ a1 ⊕CI ⊕CI I ⊃ a1 ⊕ a1 ⊕CI+I I ,
(3, 3) = (1, 1)−2,−2 ⊕ (2, 2)1,1 ⊕ (2, 1)1,−2 ⊕ (1, 2)−2,1
= (1, 1)−4 ⊕ (2, 2)2 ⊕ (2, 1)−1 ⊕ (1, 2)−1.
(9)
Tables A1–A4 in Appendix A describe the content of these MS, root by root.
The magic of the MS is that each tip {r, s} of the star, {r, s} ∈ {{0,±2}, {±1,±1}},
both in the case of e8 and of e6, can be viewed as a cubic (simple) Jordan algebra J{r,s},
over the octonions and the complex field, respectively, and each pair of opposite tips, with
respect to the center of the star, has a natural algebraic structure of a Jordan pair. The algebra
in the center of the star is the derivation algebra of the Jordan pair; when the Jordan pair is
made of a pair of Jordan algebras, its derivations also define the Lie algebra of the structure
group of the Jordan algebra itself [28–31].
2.3. The Standard Model
In this section, we relate the e8 charges to the degrees of freedom of the Standard
Model (SM) of elementary particle physics. It is not our aim to carry through a detailed
analysis; in particular, we do not consider symmetry breaking, nor the Higgs mechanism,
nor chirality and parity violation by weak interactions in the fermionic sector. We do,
however, focus on spin as an internal degree of freedom, and this will be instrumental
for the treatment of the Poincaré action on our algebra, which we will investigate in the
companion paper [1].
The first important step, after splitting the roots into colored and colorless, as in the
previous section, is to find the electromagnetic u(1)em that gives the right charges to quarks
and leptons. The choice may not be unique, although is strongly limited by the requirement
to yield the right charges, but the one we make is certainly consistent. We select the u(1)em
generated by
hγ = −i (
1
3
hα1 +
2
3
hα2 + hα3),
(10)
giving to xα, where α = ∑ λiki, the charge
qe.m.(α) := (α, qγ) = −
1
3
(λ1 + λ2 + λ3) + λ4,
(11)
where
qγ := −
1
3
α1 −
2
3
α2 − α3 = −
1
3
(k1 + k2 + k3) + k4.
(12)
The second column of Tables A1–A4 in Appendix A shows the charges of the e8 generators
xα, with α shown in the first column. In particular, Tables A3 and A4 show the charges
given to quarks and leptons.
We now select the semi-simple Lie algebra d2 ' so(4,C) ' a1⊕ a1 with roots±k5± k6
(other choices would be equivalent; thus, this does not imply any loss of generality). We
denote by ρ1, ρ2 the roots k5 − k6 and k5 + k6, respectively, by a
(1)
1
, the subalgebra of d2
Symmetry 2021, 13, 2342
9 of 25
associated to the roots±ρ1, and by a
(2)
1
, the one associated to the roots±ρ2. The non-Cartan
generators of e8 fall into a1 ⊕ a1 irreducible representations of spin (s1, s2):
(0, 0) :
xα, α = ±ki ± k j, i, j /∈ {5, 6};
(1/2, 0) :
xα, α = 12 (k± (k5 − k6)), k := ±k1 ± k2 ± k3 ± k4 ± k7 ± k8;
(0, 1/2) :
xα, α = 12 (k± (k5 + k6));
(1/2, 1/2) : xα, α = ±ki ± k5 or α = ±ki ± k6, for a fixed i /∈ {5, 6};
(1, 0) :
xα, α = ±(k5 − k6), the adjoint of a
(1)
1
(adding its Cartan);
(0, 1) :
xα, α = ±(k5 + k6), the adjoint of a
(2)
1
(adding its Cartan),
(13)
where (1, 0)⊕ (0, 1) corresponds to the six components as a rank 2 antisymmetric tensor in
four dimensions, with selfdual and antiselfdual parts (1, 0) and (0, 1), respectively. Notice
that all fermions have a half-integer spin ( 12 , 0) or (0,
1
2 ), whereas all bosons of type xα have
an integer spin.
In order to define the action of the Poincaré group in [1], we need the covering group
of rotations in the internal space. For this purpose, we select the spin (diagonal) subalgebra
su(2)spin ∈ a(1)
1 ⊕ a
(2)
1 as the compact (real) form with generators
R+ + R−,
i (R+ − R−),
i HR, where
R+ := xρ1 + xρ2 , R
− := x−ρ1 + x−ρ2 , HR :=
1
2 (hρ1 + hρ2)
(14)
The ( 12 ,
1
2 ) representation splits into a scalar and a vector under this rotation subalgebra.
The spin-1 particle within ( 12 ,
1
2 ) is the linear span of the generators xki±k5 with z-component
of spin sz := 12 (ρ1 + ρ2, ki ± k5) = (k5, ki ± k5) = ±1 and
1
2 (ε(ρ1, ki − k5)xki−k6 + ε(ρ2, ki −
k5)xki+k6) with sz = 0; the corresponding scalar is
1
2 (ε(ρ1, ki− k5)xki−k6 − ε(ρ2, ki− k5)xki+k6),
as it can be easily verified.
2.3.1. W±
Let us now consider the W± bosons. There are not many choices for them; indeed, they
must be colorless vectors with respect to su(2)spin and have electric charge ±1. The W±
bosons are therefore the generators associated to ±(k4 − k5) (within a
f
2 mentioned in
Section 2.2) and ±(k4 + k5), and the electric charge ±1 given by the presence of k4; they
change flavor to both quarks and leptons. The above analysis suggests that the extra degree
of freedom needed, say, for W+, to become massive, from the two degrees of freedom of
the massless helicity-1 state, is ε(ρ1, k4 + k6)xk4+k6 + ε(ρ2, k4 − k6)xk4−k6 , as a part of the
Higgs mechanism, which we will not discuss any further in this paper.
Remark 2. We could have made other equivalent choices for the d2 ' so(4,C) ' a1 ⊕ a1
subalgebra, which has to act non-trivially on W±: once passing to real forms, it cannot possibly
commute with the weak interaction su(2)L. For what concerns the no-go theorem by Distler–
Garibaldi, the hypothesis TOE1 of [2] cannot possibly apply, as outlined in the introduction; see
Remark 1. We also emphasize that, contrary to [2], we are dealing with the complex form of e8
because we want complex phases for the particle states.
Using the properties of the asymmetry function and the ordering of the simple roots
αi < αi+1, we obtain:
ε(ρ1, k4 − k5) = ε(ρ2, k4 − k5) = 1,
(15)
hence, the massive W± is described by three components:
W±
1
:= x±k4+k5
(sz = 1);
W±
0
:= 12 (x±k4−k6 + x±k4+k6)
(sz = 0);
W±−1 := x±k4−k5
(sz = −1).
(16)
Symmetry 2021, 13, 2342
10 of 25
Moreover, using the notation
Rx := 12 (xρ1 + xρ2 + x−ρ1 + x−ρ2);
Ry := i2 (xρ1 + xρ2 − x−ρ1 − x−ρ2);
Rz := i2 (hρ1 + hρ2),
(17)
we obtain
[
Rx, W±−1
]
= W±
0 ,
[
Rx, W±
1
]
= W±
0 ,
[
Rx, W±
0
]
= − 12
(
W±
1 + W
±
−1
)
,
[
Ry, W±−1
]
= i W±
0 ,
[
Ry, W±
1
]
= −i W±
0 ,
[
Ry, W±
0
]
= − i2
(
W±
1 −W
±
−1
)
,
(18)
and
[
Rz, W±
sz
]
= i szW±
sz , sz ∈ {1, 0,−1}.
(19)
These commutation relations correspond to the action of the rotation matrices:
Jx := 1√2
 0 −1 0
1
0
1
0 −1 0
,
Jy :=
i√
2
 0 −1 0
−1
0
1
0
1
0
,
Jz := i
 1 0
0
0 0
0
0 0 −1

(20)
regarding the vectors, v = v+e+ + vzez + v−e− ∈ R3 in the spherical basis {e+, ez, e−},
which corresponds, for angular momentum 1, to the spherical harmonic basis for the
irreducible representations of SO(3). With respect to the same vector v = vxex + vyey +
vzez ∈ R3 in the standard orthogonal basis
{
ex, ey, ez
}
, we have:
e+ = 1√2
(
ex + i ey
)
,
e− = 1√2
(
ex − i ey
)
,
v+ = 1√2
(
vx − i vy
)
,
v− = 1√2
(
vx + i vy
)
.
(21)
The transformation between (column) vectors in the two bases is represented by the unitary
matrix U:
 v+
vz
v−
 = U
 vx
vy
vz
, U :=

1√
2
0 − i√
2
0
1
0
1√
2
0
i√
2
.
(22)
The correspondence with R’s and W’s is:
Rx ↔ Jx,
Ry ↔ Jy,
Rz ↔ Jz,
W±
1 ↔ v+, W
±
−1 ↔ v−, W
±
0 ↔
1√
2
vz.
(23)
Remark 3. We have an interesting relationship between the weak and rotation generators in the
internal space (spin generators) by noticing that
[
W+
1 + W
+
−1, W
−
0
]
=
[
W−
1 + W
−
−1, W
+
0
]
= Rx
[
W+
1 −W
+
−1, i W
−
0
]
=
[
W−
1 −W
−
−1, i W
+
0
]
= Ry
(24)
and, consequently, the relation with Rz = [Rx, Ry].
2.3.2. Z0
We associate the Z0 boson with spin sz = 0, denoted by Z00 , to the vector orthogonal
to qγ in the plane of (k4 − k5) and qγ; hence, it is, up to a scalar, the Cartan generator
Z00 =
1
4 (hα1 + 2hα2 + 3hα3 + 4hα4). It interacts with left-handed neutrinos and right-handed
antineutrinos, contrary to the photon; it does not allow for flavor-changing neutral currents.
Notice that the generator of hypercharge u(1)Y of the standard model is, in this, the
setting compact Cartan generator i hY, where hY := − 16 (2hα1 + 4hα2 + 6hα3 + 3hα4). The
Weinberg angle θW is the angle between the axis representing the photon and the axis
Symmetry 2021, 13, 2342
11 of 25
representing the hypercharge; therefore, θW = π/2− φ, where φ is the angle between qγ
and k4 − k5, and we obtain sin2θW = 3/8.
Since α1 + 2α2 + 3α3 + 4α4 = k1 + k2 + k3 + k4 − 4k5, 14 (α1 + 2α2 + 3α3 + 4α4, x±ρ1,2)
= −(k5,±ρ1) = −(k5,±ρ2) = ∓1; hence, we have the following commutation relations:
[xρ1 + xρ2 + x−ρ1 + x−ρ2 , Z
0
0 ] = (xρ1 + xρ2 − x−ρ1 − x−ρ2);
[
xρ1 + xρ2 − x−ρ1 − x−ρ2 , Z00
]
= (xρ1 + xρ2 + x−ρ1 + x−ρ2),
(25)
that is
[
Rx, Z00
]
= −i Ry,
[
Ry, Z00
]
= i Rx,
[
Rz, Z00
]
= 0.
(26)
We want Z0, as a spin-1 particle, to obey the same commutation relations with the
rotation generators as W±. We can define the spin ±1 components of Z0 this way by a
comparison with the last commutator in each row of (18):
Z01 = −
[
Rx, Z00
]
+ i
[
Ry, Z00
]
= −Rx + i Ry = −R+;
Z0−1 = −
[
Rx, Z00
]
− i
[
Ry, Z00
]
= Rx + i Ry = R−.
(27)
(see (14) for the definition of R±).
By looking at Table A3 in Appendix A, we notice that Z01 interacts, for instance, with ν
′
e
with spin − 12 to give ν′e with spin
1
2 , and, similarly, for other leptons and for quarks.
In particular, there are no flavor-changing neutral currents.
2.3.3. The Tables at a Glance
From the Tables A1–A4 in Appendix A, one can deduce all the standard model
charges (in particular, we have denoted with a prime possible mixings in Tables A3 and A4).
We have:
(SM.1)
The color charges are denoted by the pair {rc, sc}, and one can associate colors
to them—say, blue = {1, 1}, green = {−1, 1} and red = {0,−2}—and, similarly,
for the anti-colors;
(SM.2)
The quarks are the fermions in Table A1 with a certain color; they come in three
color families, and anti-quarks have anti-colors and opposite electric charges
− 23 ,
1
3 with respect to quarks;
(SM.3)
The gluons are the generators of ac2, change color to the quarks on which they
act, as on a 3 or3̄ representation, and their electric charge is 0;
(SM.4)
The leptons are in the center of the MS in Table A1 and are the fermions in
Table A2;
(SM.5)
The leptons have an integer electric charge in {−1, 0, 1};
(SM.6)
There are four flavor families; we have used the notation χ, νχ for the fourth
lepton family and T, B for the fourth quark family;
(SM.7)
The fourth column of Tables A3 and A4 shows the component of spin along the
axis specified by the spin generator Rz := i2 (hρ1 + hρ2). Obviously, a rotation
by 2π of quarks and leptons changes their sign, regardless of whether it leaves
vector bosons invariant.
The consequences of this classification, with respect to the Poincaré action on gu, will
be discussed in the companion paper [1].
3. The Kac–Moody Algebras e9, e10, e12, de12
Let g(A) denote the Kac–Moody algebra associated to the n × n Cartan matrix A,
with Cartan subalgebra h. For all algebras in this paper, A is symmetric; its entries are
denoted by aij. We denote the Chevalley generators by Ei, associated to the simple root αi,
and by Fi, associated to the root −αi. Let n+ (resp. n−) denote the subalgebra of g(A)
Symmetry 2021, 13, 2342
12 of 25
generated by {E1, . . . , En} (resp.{F1, . . . , Fn}). By Theorem 1.2 (a), (e) in [27], the following
triangular decomposition holds:
g(A) = n+ ⊕ h⊕ n−
(direct sum of vector spaces).
(28)
Note that, for a root α > 0 (resp. α < 0), we have α ∈ h∗, the dual of h, and the vector
space gα = {x ∈ g(A) | [h, x] = α(h)x ∀h ∈ h} is the linear span of the elements of the
form [. . . [[Ei1 , Ei2 ], Ei3 ] . . . Eis ] (resp. [. . . [[Fi1 , Fi2 ], Fi3 ] . . . Fis ]), such that αi1 + . . . + αis = α
(resp. = −α). The multiplicity mα of a root α is defined as mα := dim gα (mαi = m−αi = 1
for each simple root αi).
Kac–Moody algebras [27,32] can be tackled in terms of simple roots and their (ex-
tended) Dynkin diagram, or, equivalently, their Cartan matrix, without any reference to
root coordinates. Some physical features or interpretations may, however, be more explicit
when roots are expressed on an orthonormal basis rather than on a simple root basis. This
is the case of this paper, in which some root coordinates, except for the case of e8, are
interpreted as momentum coordinates. We recall that the metric is Euclidean for e8, but
Lorentzian in the case of e9, e10, e12, de12.
Our notation for the simple roots is shown in the Dynkin diagram of e12:
e
e
e
e
e
e
e
e
e
e
e
e
α-1
α0′′
α0′
α0
α1
α2
α3
α4
α5
α7
α8
α6
(29)
Analogous diagrams are those of e10 (without α-1, α0′′ ), of e9 (without α-1, α0′′ , α0′ ) and of
e8 (without α-1, α0′′ , α0′ , α0).
3.1. The Simple Roots of e9 and e10
We introduce the following set of simple roots α-1, α0, α1, . . . , α8 of e10 in terms of the
basis vectors k-1, k0, k1, . . . , k8 spanning the Lorentzian space R9,1, with (k-1, k-1) = −1 and
(ki, ki) = 1, for 0 ≤ i ≤ 8:
α−1 =
1
2 (k−1 − 3k0);
α0 =
1
2 (k−1 + k0)− k1 + k8;
α1 =
k1 − k2;
α2 =
k2 − k3;
α3 =
k3 − k4;
α4 =
k4 − k5;
α5 =
k5 − k6;
α6 =
k6 − k7;
α7 =
k6 + k7;
α8 = − 12 (k1 + k2 + k3 + k4 + k5 + k6 + k7 + k8).
(30)
All of these roots have norm 2, with respect to the scalar product (·, ·), and the correspond-
ing Cartan matrix is the Gram matrix of the e10 even unimodular Lorentzian lattice II9,1
made of all the vectors in R9,1, whose components are all in Z or all in Z+ 12 and have an
integer scalar product with 12 (k-1 + k0 + k1 + k2 + k3 + k4 + k5 + k6 + k7 + k8), as can be
easily checked.
The affine Kac–Moody algebra e9 is obtained by eliminating the root α-1. Notice that
α0 = α9 + δ ,
{
δ := 12 (k-1 + k0),
(hence (δ, δ) = 0);
α9 := −(2α1 + 3α2 + 4α3 + 5α4 + 6α5 + 3α6 + 4α7 + 2α8),
(31)
where α9 = −k1 + k8 is the lowest height root of e8 and δ is a light-like vector.
Symmetry 2021, 13, 2342
13 of 25
3.2. e9
The Cartan subalgebra of e9 is the span of 10 generators, containing two new ones
with respect to e8. We write
h = span{K, d, hi | i = 1, . . . , 8},
(32)
where
K
:= hδ, δ := 12 (k−1 + k0);
d
:= hρ, ρ := −k−1 + k0,
(33)
and K is a central element.
Let Φ8 be the root system of e8, h be its Cartan subalgebra, hα (resp. eα) be a Cartan
(resp. non-Cartan) generator associated to the root α, hi for 1 ≤ i ≤ 8 be a Cartan generator
associated to the simple root αi and X and Y be either Cartan or non-Cartan generators of
e8. It is shown by Kac [27] that:
1.
The root system Φ9 of e9 is
Φ9 := {α+ mδ | α ∈ Φ8 , m ∈ Z} ∪ {mδ | m ∈ Z\0};
(34)
2.
e8 is determined by the following commutation relations:
[h, h′]
= 0
if h, h′ ∈ h;
[h, Eα]
= (h∗, α)Eα
if h ∈ h, α ∈ Φ8;
[Eα, E−α] = −hα
if α ∈ Φ8;
[
Eα, Eβ
]
= 0
if α, β ∈ Φ8, α+ β /∈ Φ8 ∪ {0};
[
Eα, Eβ
]
=
ε(α, β)Eα+β
if α, β, α+ β ∈ Φ8,
(35)
where ε is Kac’s asymmetry function, see Section 2.1;
3.
The commutation relations in e9 are the same as those for the central extended loop
algebra of e8 plus derivations:
[tm ⊗ X⊕ λK⊕ µ d, tn ⊗Y⊕ λ1K⊕ ν d] =
(tm+n ⊗ [X, Y]−mνtm ⊗ X + nµtn ⊗Y)⊕mδm,−n(X|Y) K
(36)
with the following correspondence
tm ⊗ Eα ↔ xα+mδ;
tm ⊗ Hα ↔ xαmδ if m
6= 0;
t0 ⊗ Hα ↔ hα;
K, d
↔ hδ, hρ, (see Equation (33))
(37)
and with the invariant non-degenerate symmetric bilinear form (·|·) defined by:
(X|Y) :=

(α, β)
if X = Hα, Y = Hβ;
0
if X = Hα, Y = Eβ;
−δα,−β
if X = Eα, Y = Eβ.
(38)
For α, β ∈ Φ8 (roots of e8) and the letter h referring to a Cartan generator, the commu-
tation relations, with no reference to the loop algebra, are:
Symmetry 2021, 13, 2342
14 of 25
[h, h′] = 0
[
hα, x
β
mδ
]
= 0
m
6= 0
[
hδ, xαmδ
]
= 0
m
6= 0
[
hρ, xαmδ
]
= mxαmδ
m
6= 0
[
hβ, xα+mδ
]
= (β, α)xα+mδ
[hδ, xα+mδ] = 0
[
hρ, xα+mδ
]
= mxα+mδ
[
xαmδ, x
β
nδ
]
= mδm,−n(α, β)hδ m, n
6= 0
[
xαmδ, xβ+nδ
]
= (α, β)xβ+(m+n)δ m
6= 0
(39)
[
xα+mδ, xβ+nδ
]
=

0
if α+ β
6∈ Φ8 ∪ {0}
ε(α, β)xα+β+(m+n)δ
if α+ β ∈ Φ8
−xα
(m+n)δ
if α+ β = 0 and m + n
6= 0
−hα+mδ
if α+ β = 0 and m + n = 0
(40)
Remark 4. The commutation relations of e9 are essentially determined by those of e8, whose main
ingredient for explicit calculations is the asymmetry function.
Remark 5. The second correspondence in (37) shows why the so called imaginary roots mδ are eight-
fold degenerate: the space of generators associated to each root mδ is indeed an eight-dimensional
space isomorphic to the span of {hi, i = 1, . . . , 8}, namely the Cartan subalgebra of e8.
3.3. e9 in a 1 + 1 Dimensional Toy Model
The explicit construction presented here, using a realization of the roots in terms of
the orthonormal basis {ki , i = −1, 0, 1, . . . , 8} of R9,1, suggests letting the coordinates
k1, . . . , k8 relate to charge/spin degrees of freedom, and interpreting the coordinates k-1, k0
as 2-momentum coordinates with a Lorentzian signature.
A crucial step in our model, which describes a universe that expands from an ini-
tial quantum state, is to restrict the particles forming that state, and hence their inter-
actions, to lie in the subalgebra n+ of the triangular decomposition Equation (28) of e9
(the reason will be explained in item (TM.3) below). The restriction to n+ has the follow-
ing consequences:
(TM.1)
The only commutation relations within (39) and (40) occurring in n+ are, for
α, β ∈ Φ8:
[
xαmδ, x
β
nδ
]
= 0
m, n
6= 0
[
xαmδ, xβ+nδ
]
= (α, β)xβ+(m+n)δ
m > 0, n ≥ 0
[
xα+mδ, xβ+nδ
]
=

0
ε(α, β)xα+β+(m+n)δ
−xα
(m+n)δ
if α+ β
6∈ Φ8 ∪ {0}
if α+ β ∈ Φ8
if α+ β = 0
(41)
We remark that α+ β = 0 in the last commutation relation implies that either
α or β is negative, hence m + n
6= 0, being m, n ≥ 0 and, in particular, m > 0 in
xα+mδ whenever α is a negative root of e8;
(TM.2)
The roots involved in the interactions are not only real; for instance, α0,−α9 are
positive roots and (α0,−α9) = −2. Therefore, by Proposition 5.1 of [27], at least
α0 + (−α9) = δ and α0 + 2(−α9) are roots; the outgoing generator xα9
δ of the
interaction between xα0 = xα9+δ and x−α9 is similar to the Cartan generator hα9 ,
except that it carries a momentum δ = 12 (k-1 + k0). This yields the interesting
consideration that neutral radiation fields, like the photon, are not associated to Cartan
elements of infinite dimensional Kac–Moody algebras, but to imaginary roots, a feature
that is not present in Yang–Mills theories. It is also worth noticing that the second
Symmetry 2021, 13, 2342
15 of 25
equation in (41) implies that the neutral radiation field keeps memory of the
particle–antiparticle pair that produced it, represented in the equation by the
root α;
(TM.3)
The fact that all particles are in n+ ensures that their energy is always positive,
even though they may be related to both positive and negative roots of e8,
as revealed by the fact that α9 is the negative root of the lowest height. In other
words, we obtain both particles and antiparticles, and all of them do have
positive energy;
(TM.4)
The momenta given to each particle by the interactions are light-like. Energy mo-
mentum is conserved because the outgoing particle in an elementary interaction
is associated to a root, which is the sum of the roots of the incoming particles.
All particles are massless, since momenta add up in the unique spatial direction;
(TM.5)
We give fermionic particles helicity 1/2.
So far, everything runs smooth and seems physically plausible. However:
(TM.6)
The initial quantum state of the two-dimensional toy model under consideration
is to be a superposition of states with momenta in opposite space directions and
opposite helicity;
(TM.7)
Sinceδ̃ := 12 (k-1 − k0) is not a root of e9, we need to introduce the auxiliary roots
α+ mδ̃
(42)
yielding the needed superposition of momenta. We will use the notation:
p̃
:= mδ̃ for p = mδ;
x̃α+p
:= xα+p + ηαxα+p̃, where η4α = 1;
(43)
x̃αp
:= xαp + ηαx
α
p̃,
so that {xα+p → ηαxα+p̃, xαp → xα
p̃} is an isomorphism e9 →ẽ9 and xα → ηαxα is
an e8 involution, up to a sign. The coefficient ηα ∈ {±1,±i } has been introduced
to have the freedom of varying it depending on the spin of the generator related
to α.
The commutation relations (41) become, for α, β ∈ Φ8 and p1, p2 linear combina-
tion with positive integer coefficients of δ,δ̃:
[
xαp1 , x
β
p2
]
= 0
p1, p2
6= 0
[
xαp1 , xβ+p2
]
= (α, β)xβ+p1+p2
p1
6= 0
[
xα+p1 , xβ+p2
]
=

0
ε(α, β)xα+β+p1+p2
−xαp1+p2
if α+ β
6∈ Φ8 ∪ {0}
if α+ β ∈ Φ8
if α+ β = 0
(44)
α+ β = 0 in the last commutation relation implies p1 + p2
6= 0, as remarked in
item (TM.1). Moreover, (p1 + p2)2 ≥ 0, and we still have positive energy associ-
ated to all particles. It is no longer true, however, that particles are necessarily
massless, as we immediately realize by the fact that δ+δ̃ = k−1 represents a
mass at rest. We also notice that the product is to be antisymmetric; therefore,
x−α
p = −xαp and
[
xαp1 , xβ+p2
]
= −
[
xβ+p2 , x
α
p1
]
(45)
Moreover, for consistency
xα+β
p = xαp + x
β
p
(46)
We will prove in the forthcoming paper [1], as a particular case of a more general
statement, that the algebra so defined is a Lie algebra;
Symmetry 2021, 13, 2342
16 of 25
(TM.8)
We lack two spatial dimensions. This suggests a further extension to e12 or de12,
or to analogous Borcherds (or generalized Kac–Moody) algebras, as we will
investigate in the next sections;
(TM.9)
Our toy model still lacks three features, which urges a further extension of the
algebra (investigated in the companion paper [1]):
(a)
Locality, i.e., spacetime-related multiplication rules that immerse the alge-
bra into a vertex-type algebra;
(b)
Space expansion within the vertex algebra;
(c)
Pauli exclusion principle that, as we will see, requires an extension to
Lie superalgebra.
The above considerations imply the fact that the extension of e8 to Kac–Moody, or, even
beyond, to generalized Kac–Moody (Borcherds) algebras, is very appealing to particle
physics, and not only to two-dimensional conformal field theory [33].
3.4. e12 and de12
The Dynkin diagram of e12 is shown in (29). We use the same indices for the simple
roots α-1, α0′′ , α0′ , α0, α1, . . . , α8 and the orthonormal basis vectors k-1, k0′′ , k0′ , k0, k1, . . . , k8
of the Lorentzian space R11,1, with (ki, ki) = 1, for i = 0′′, 0′, 0, 1, . . . , 8, and (k-1, k-1) = −1.
Here, it is worth presenting a different choice for the set of simple roots, with respect
to (30), in which the e9 simple roots are all fermionic:
α−1 = −
√
2
2 k0′′ +
√
6
2 k0′ +
√
3δ;
α0′′ =
√
2k0′′ + δ;
α0′ =
√
2k0′ − 2k0 + 2δ;
α0 =
1
2 (−k1 + k2 + k3 − k4 − k5 − k6 − k7 − k8) + δ;
α1 =
1
2 (k1 − k2 − k3 − k4 + k5 + k6 + k7 − k8);
α2 =
1
2 (k1 − k2 + k3 + k4 − k5 − k6 − k7 + k8);
α3 =
1
2 (−k1 + k2 − k3 + k4 − k5 + k6 + k7 − k8);
α4 =
1
2 (k1 + k2 − k3 − k4 + k5 − k6 − k7 + k8);
α5 =
1
2 (−k1 − k2 + k3 + k4 + k5 − k6 + k7 − k8);
α6 =
1
2 (k1 + k2 + k3 − k4 − k5 + k6 + k7 + k8);
α7 =
1
2 (−k1 − k2 − k3 − k4 − k5 + k6 − k7 + k8);
α8 =
1
2 (k1 + k2 + k3 + k4 + k5 + k6 − k7 − k8),
(47)
where δ = 12 (k0 + k-1).
The corresponding Cartan matrix is the Gram matrix of the e12 lattice in R11,1, which
is not unimodular. We interpret the coordinates (k0′′ , k0′ , k0, k-1) as four-momentum coordi-
nates with a Lorentzian signature.
Similar arguments hold for the Kac–Moody algebra de12, whose Dynkin diagram is:
e
e
e
e
e
e
e
e
e
e
e
e
α0′′ α0′
α0
α1
α2
α3
α4
α5
α7
α8
α6
α-1
(48)
This is the extension of e8 through the orthogonal Lie algebra d4.
Symmetry 2021, 13, 2342
17 of 25
A possible set of simple roots in the orthonormal basis of the Lorentzian space R11,1 is:
α−1 =
√
2k0′ + 3δ;
α0′′ =
√
2k0′′ + δ;
α0′ =
√
2k0′ − 2k0 + 2δ;
α0 =
1
2 (−k1 + k2 + k3 − k4 − k5 − k6 − k7 − k8) + δ;
α1 =
1
2 (k1 − k2 − k3 − k4 + k5 + k6 + k7 − k8);
α2 =
1
2 (k1 − k2 + k3 + k4 − k5 − k6 − k7 + k8);
α3 =
1
2 (−k1 + k2 − k3 + k4 − k5 + k6 + k7 − k8);
α4 =
1
2 (k1 + k2 − k3 − k4 + k5 − k6 − k7 + k8);
α5 =
1
2 (−k1 − k2 + k3 + k4 + k5 − k6 + k7 − k8);
α6 =
1
2 (k1 + k2 + k3 − k4 − k5 + k6 + k7 + k8);
α7 =
1
2 (−k1 − k2 − k3 − k4 − k5 + k6 − k7 + k8);
α8 =
1
2 (k1 + k2 + k3 + k4 + k5 + k6 − k7 − k8),
(49)
where δ := 12 (k0 + k-1). One can realize at a glance that this is the same set of simple roots of
e12, except for the root α-1. We think that the possibility of discriminating between e12 and
de12 on physical grounds can only arise when performing explicit computer calculations,
which, for the case of de12, should undergo major simplifications due to the presence of
only one irrational number,
√
2.
4. Beyond Kac–Moody
Several difficulties in proceeding with our program arise with Kac–Moody algebras:
(P.1)
As already mentioned above, the presence of irrational numbers in the definition
of momentum variables can be very annoying in computer calculations, and it is
also quite unnatural in an algebra based on integer numbers, both in the roots and
in the structure constants;
(P.2)
The need of roots, as in item (TM.7) for e9, with opposite helicity and opposite
signs of the three-momentum components k0′′ , k0′ , k0, complicates the algebra to
a large extent, since they cannot, in pairs, be roots of the Kac–Moody algebras. In the
case of e9, this problem can be overcome by enlarging the explicit commutation
relations to consistently include the new roots, but, in the case of e12 or de12, one
needs to further modify the Serre relations, which does not seem to be an easy task
to us;
(P.3)
However, the most important issue comes from physics: in e12 and de12, three
simple roots (namely, α-1, α0′′ and α0′ ) have tachyon-like momenta due to their
positive norm. The interpretation of such tachyonic momenta, as well as the
investigation of their impact on the interactions among charged particles, is beyond
the scope of the present paper; computer calculations, starting from an initial state,
may reveal the scenario that tachyonic simple roots may yield to.
It is our opinion that these are good motivations for focusing our investigation on
generalized Kac–Moody (Borcherds) algebras, where two of the three difficulties listed
above disappear.
5. BorcherdsB12
Borcherds algebras are a generalization of Kac–Moody algebras obtained by releasing
the condition on the diagonal elements of the Cartan matrix, which are then allowed to
be non-positive, as well as by restricting the Serre relations to the generators associated to
positive norm simple roots [34,35].
A generalized Kac–Moody (or Borcherds) algebra B is constructed as follows.
Symmetry 2021, 13, 2342
18 of 25
Let H be a real vector space with a symmetric bilinear inner product (·, ·), and with
elements hi indexed by a countable set I , such that (hi, hj) ≤ 0 if i
6= j and 2(hi, hj)/(hi, hi)
is an integer if (hi, hi) is positive. The matrix A with entries aij := (hi, hj) is called the
symmetrized Cartan matrix of B.
The generalized Kac–Moody (or Borcherds) algebra B associated to A is defined
to be the Lie algebra generated by H and elements ei and fi, for i ∈ I , with the follow-
ing relations:
1.
The (injective) image of H in B is commutative;
2.
If h is in H, then [h, ei] = (h, hi)ei and [h, fi] = −(h, hi) fi;
3.
[ei, f j] = δijhi;
4.
If aii > 0 and i
6= j, then ad(ei)n ej = ad( fi)n f j = 0, where n = 1− 2aij/aii;
5.
If aij = 0, then [ei, ej] = [ fi, f j] = 0.
If aii > 0 for all i ∈ I , then B is the Kac–Moody algebra with Cartan matrix A.
In general, B has almost all of the properties of a Kac–Moody algebra, the only major
difference being that B is allowed to have imaginary simple roots.
The root lattice L is the free Abelian group generated by elements αi for i ∈ I , called
simple roots, and L has a real-valued bilinear form defined by (αi, αj) = aij . The Lie algebra
B is then graded by L with H in degree 0, ei (resp fi) in degree αi (resp. −αi). A root is a
non-zero element α of L such that there are elements of B of degree α. A root r is called
real if (r, r) > 0; otherwise, it is called imaginary. A root r is positive if it is a sum of simple
roots, and negative if −r is positive. Notice that every root is either positive or negative, [34].
We build the following symmetrized Cartan matrix for a Borcherds algebra of rank 12,
which we denote by B12:

−1 −1 −1 −1
0
0
0
0
0
0
0
0
−1
0 −1 −1
0
0
0
0
0
0
0
0
−1 −1
0 −1
0
0
0
0
0
0
0
0
−1 −1 −1
2 −1
0
0
0
0
0
0
0
0
0
0 −1
2 −1
0
0
0
0
0
0
0
0
0
0 −1
2 −1
0
0
0
0
0
0
0
0
0
0 −1
2 −1
0
0
0
0
0
0
0
0
0
0 −1
2 −1
0
0
0
0
0
0
0
0
0
0 −1
2 −1 −1
0
0
0
0
0
0
0
0
0 −1
2
0
0
0
0
0
0
0
0
0
0 −1
0
2 −1
0
0
0
0
0
0
0
0
0
0 −1
2

.
(50)
Notice that, for δ := α0 + 2α1 + 3α2 + 4α3 + 5α4 + 6α5 + 3α6 + 4α7 + 2α8, a four-
momentum vector can be written as
p := p0α-1 + p1(α0′′ − α-1) + p2(α0′ − α-1) + p3(δ− α-1).
(51)
Using the Cartan Matrix (50), we indeed obtain:
(α-1, α-1) = −1;
(α0′′ − α-1, α0′′ − α-1) = (α0′ − α-1, α0′ − α-1) = (δ− α-1, δ− α-1) = 1;
(α-1, α0′′ − α-1) = (α-1, α0′ − α-1) = (α-1, δ− α-1) = 0;
(α0′′ − α-1, α0′ − α-1) = (α0′′ − α-1, δ− α-1) = (α0′ − α-1, δ− α-1) = 0,
(52)
hence the Lorentzian scalar product:
(p, p′) = −p0 p′0 + p1 p′1 + p2 p′2 + p3 p′3.
(53)
Symmetry 2021, 13, 2342
19 of 25
Let us restrict to positive roots r = ∑I λiαi, I = {−1, 0′′, 0′, 0, . . . , 8}, with λi ∈ N∪ {0},
and let us denote by B+ the corresponding subalgebra of B12. The physical motivation for
restricting to B+ is that, given a positive root r = ∑I λiαi, its four-momentum is
p = (p0, p1, p2, p3) = (λ−1 + λ0′′ + λ0′ + λ0,λ0′′,λ0′,λ0),
(54)
with λ−1,λ0′′ ,λ0′ ,λ0 ≥ 0, implying m2
:= −p2 ≥ 0, namely p is either light-like or
time-like. In particular:
p2 = −(λ2−1 + 2λ−1 ∑i
6=−1 λi +∑i
6=j, i,j
6=−1 λiλj) , i, j ∈ {−1, 0′′, 0′, 0}

= 0
if λ−1 = 0 and at most one λi
6= 0 , i
6= −1
= −1
if λ−1 = 1 and all λi = 0 , i
6= −1
≤ −2
otherwise
(55)
Remark 6. Notice that the mass of a particle cannot be arbitrarily small, since there is a lower limit
m ≥ 1.
For r = ∑I λiαi, I = {−1, 0′′, 0′, 0, ..., 8}, with λi ∈ N∪ {0}, we introduce the notation
r = α+ p p := λ−1α−1 + λ0′′α0′′ + λ0′α0′ + λ0δ (see (51) and (54))
α := λ0α9 + λ1α1 + ... + λ8α8
= (λ1 − 2λ0)α1 + (λ2 − 3λ0)α2 + (λ3 − 4λ0)α3
+ (λ4 − 5λ0)α4 + (λ5 − 6λ0)α5 + (λ6 − 3λ0)α6
+ (λ7 − 4λ0)α7 + (λ8 − 2λ0)α8
(56)
Thus, α is in the lattice Le8 of e8, and a precise physical meaning is assigned to positive
real and imaginary roots when α ∈ Le8\{0}:
Proposition 3. A generator in B+, associated to a positive root r = α+ p, with α ∈ Le8\{0}
and momentum p
6= 0, is massive if and only if α+ p is an imaginary root; it is massless if and
only if r is real, in which case, it is a positive real root of e9 ⊂ B12.
Proof. The proof consists of the following steps:
1.
From Proposition 2.1. of [34], it holds that every positive root r = α+ p is conjugate
under the Weyl group to a root r0 = α′ + p′, such that either r0 is a simple real root αi,
i ∈ {0, 1, . . . , 8}, or it is a positive root in the Weyl chamber (namely (r0, αi) ≤ 0 for all
simple roots αi);
2.
Since r and r0 are conjugate under the Weyl group, then (r, r) = (r0, r0);
3.
If r0 is a real simple root, then it is a root of e9 and p′
2 = 0; r0 is real and so is r. Since
the Weyl group is generated by the reflections ρ− (ρ, αi)αi, where the αi simple roots
are real, hence αi ∈ e9 ⊂ B12, it coincides with the Weyl group of e9. By applying to
r0 Weyl reflections, we stay within e9, since every Kac–Moody algebra is invariant
under the Weyl group; therefore, r is a real root of e9, namely r = α+ mδ, α ∈ e8 and
mδ is light-like;
4.
If r0 is in the Weyl chamber, then (r, r) = (r0, r0) = ∑ λi(r0, αi) ≤ 0, i ∈ I , since all λi
are positive, r0 being a positive root; thus, r is imaginary;
5.
Since (r, r) = α2 + p2 ≤ 0 with α2 ≥ 2, then m2 ≥ α2 ≥ 2, and the particle associated
to r is massive.
Symmetry 2021, 13, 2342
20 of 25
Remark 7. In the massive case, the lower limit of the mass grows with the norm of α:
if
α ∈ Le8\{0} is not a root of Φ8, then the mass is certainly bigger than the lower mass a par-
ticle corresponding to a root of Φ8 may have. We also notice that charged massless particles (α
6= 0
in the root α+ p) are quite peculiar, since their momentum can only be in one direction. The photon
is not in this class, since it has α = 0, but the (non-virtual) gluons are. A non-virtual photon can
be produced in a decay process [1].
Remark 8. We emphasize that two of the three problems listed in Section 4 about Kac–Moody
algebras vanish in the Borcherds algebra B12. These are obviously (P.1) and (P.3). However, (P.2)
still remains [1].
6. Conclusions
In this paper, the first of a series of two papers with the same title, we have described
the basic principles of a model of quantum gravity at the early stages of the universe.
We have explained how spacetime is generated from an initial state and how it expands
and is driven by interactions in a purely algebraic context. We have investigated the
mathematical structures that may suit our purpose: they are rank-12 infinite-dimensional
algebras, extending e8 and including 4-momenta. We have also discussed why a celebrated
no-go theorem on e8 does not apply in our settings.
The companion paper [1], based on the treatment and considerations of this paper,
will focus on a particular rank-12 algebra gu in order to build a model for quantum gravity.
In particular, it will turn gu into a Lie superalgebra in order to fulfill the Pauli exclu-
sion principle without producing superpartners. It will also discuss scattering processes
and decays.
The algebra gu is based on a simplified version of a Borcherds algebra, where based
on means that we still have to enlarge the algebra with roots that take into account the
coupling of four momenta with those having three momenta of the opposite sign and same
energy (see (P.2) of Section 4). Nothing would prevent us from starting from a Borcherds
algebra, but we would not have explicit commutation relations, as we do in [1].
Author Contributions: Conceptualization, P.T., A.M., M.R. and K.I.; methodology, P.T. and A.M.;
formal analysis, P.T., A.M. and M.R.; investigation, P.T.; writing—original draft preparation, P.T.;
writing—review and editing, P.T., A.M. and M.R.; supervision, K.I. All authors have read and agreed
to the published version of the manuscript.
Funding: This research received no external funding.
Conflicts of Interest: The authors declare no conflict of interest.
Appendix A
In this appendix, we show the charge content of the e8 part of the algebra, root by
root. Table A1 show all of the roots of e8, grouped by the points of the Magic Star obtained
by projecting e8 on the plane of color ac2, see Section 2.2. The leptons, being colorless, lie
in the center of this Magic Star, forming the e6 roots of Table A2. In addition, the roots of
this e6 lie on a Magic Star (of a smaller dimension) once projected on the plane of flavor a
f
2 .
The families of leptons and quarks are shown in Tables A3 and A4, together with their spin.
For the quarks sector, we only consider one color, which we name blue; similar tables for
red and green colors are trivially obtained by exchanging indices.
Symmetry 2021, 13, 2342
21 of 25
Table A1. The Magic Star of e8; qe.m.(α) = (α,− 13 (k1 + k2 + k3) + k4);
rc(α) = (α, k1 − k2),
sc(α) = (α, k1 + k2 − 2k3).
Roots
qe.m.
#
{rc, sc}
±(k1 − k2)
0
2
±{2, 0}
±(k2 − k3)
0
2
±{−1, 3}
±(k1 − k3)
0
2
±{1, 3}
±ki ± k j
5 ≤ i < j ≤ 8
0
24
k4 ± ki
5 ≤ i ≤ 8
1
8
−k4 ± ki
5 ≤ i ≤ 7
−1
8
{0, 0}
1
2 (±(k1 + k2 + k3 + k4)± . . .± k8)
even # of +
0
16
1
2 (−(k1 + k2 + k3) + k4 ± . . .± k8)
even # of +
1
8
1
2 (+(k1 + k2 + k3)− k4 ± . . .± k8)
even # of +
−1
8
−k2 − k3 , k1 + k4
2/3
2
k1 − k4
−4/3
1
k1 ± ki
i = 5, . . . , 8
−1/3
8
{1, 1}
1
2 (k1 − k2 − k3 + k4 ± . . .± k8)
even # of +
2/3
8
1
2 (k1 − k2 − k3 − k4 ± . . .± k8)
even # of +
-1/3
8
+k2 + k3 , −k1 − k4
−2/3
2
−k1 + k4
4/3
1
−k1 ± ki
i = 5, . . . , 8
1/3
8
{−1,−1}
1
2 (−k1 + k2 + k3 − k4 ± . . .± k8)
even # of +
−2/3
8
1
2 (−k1 + k2 + k3 + k4 ± . . .± k8)
even # of +
1/3
8
−k1 − k3 , k2 + k4
2/3
2
k2 − k4
−4/3
1
k2 ± ki
i = 5, . . . , 8
−1/3
8
{−1, 1}
1
2 (−k1 + k2 − k3 + k4 ± . . .± k8)
even # of +
2/3
8
1
2 (−k1 + k2 − k3 − k4 ± . . .± k8)
even # of +
−1/3
8
k1 + k3 , −k2 − k4
−2/3
2
−k2 + k4
4/3
1
−k2 ± ki
i = 5, . . . , 8
1/3
8
{1,−1}
1
2 (k1 − k2 + k3 − k4 ± . . .± k8)
even # of +
-2/3
8
1
2 (k1 − k2 + k3 + k4 ± . . .± k8)
even # of +
1/3
8
−k1 − k2 , k3 + k4
2/3
2
k3 − k4
−4/3
1
k3 ± ki
i = 5, . . . , 8
−1/3
8
{0,−2}
1
2 (−k1 − k2 + k3 + k4 ± . . .± k8)
even # of +
2/3
8
1
2 (−k1 − k2 + k3 − k4 ± . . .± k8)
even # of +
−1/3
8
k1 + k2 , −k3 − k4
−2/3
2
−k3 + k4
4/3
1
−k3 ± ki
i = 5, . . . , 8
1/3
8
{0, 2}
1
2 (k1 + k2 − k3 − k4 ± . . .± k8)
even # of +
-2/3
8
1
2 (k1 + k2 − k3 + k4 ± . . .± k8)
even # of +
1/3
8
Symmetry 2021, 13, 2342
22 of 25
Table A2. The Magic Star of e6; qe.m.(α) = (α,− 13 (k1 + k2 + k3) + k4);
r f (α) = (α, k4 − k5),
s f (α) = (α, k4 + k5 − 2k6).
Roots
qe.m.
# of Roots
{r f , s f}
±(k4 − k5)
±1
2
±{2, 0}
±(k5 − k6)
0
2
±{−1, 3}
±(k4 − k6)
±1
2
±{1, 3}
±k7 ± k8
0
4
1
2 (±(k1 + k2 + k3 + k4 + k5 + k6)± k7 ± k8)
even # of +
0
4
{0, 0}
1
2 (−k1 − k2 − k3 + k4 + k5 + k6 ± k7 ± k8)
even # of +
1
2
1
2 (k1 + k2 + k3 − k4 − k5 − k6 ± k7 ± k8)
even # of +
−1
2
−k5 − k6
0
1
k4 ± ki
i = 7, 8
1
4
{1, 1}
1
2 (k1 + k2 + k3 + k4 − k5 − k6 ± k7 ± k8)
even # of +
0
2
1
2 (−k1 − k2 − k3 + k4 − k5 − k6 ± k7 ± k8)
even # of +
1
2
k5 + k6
0
1
−k4 ± ki
i = 7, 8
−1
4
{−1,−1}
1
2 (k1 + k2 + k3 − k4 + k5 + k6 ± k7 ± k8)
even # of +
−1
2
1
2 (−k1 − k2 − k3 − k4 + k5 + k6 ± k7 ± k8)
even # of +
0
2
−k4 − k6
−1
1
k5 ± ki
i = 7, 8
0
4
{−1, 1}
1
2 (k1 + k2 + k3 − k4 + k5 − k6 ± k7 ± k8)
even # of +
−1
2
1
2 (−k1 − k2 − k3 − k4 + k5 − k6 ± k7 ± k8)
even # of +
0
2
k4 + k6
1
1
−k5 ± ki
i = 7, 8
0
4
{1,−1}
1
2 (k1 + k2 + k3 + k4 − k5 + k6 ± k7 ± k8)
even # of +
0
2
1
2 (−k1 − k2 − k3 + k4 − k5 + k6 ± k7 ± k8)
even # of +
1
2
−k4 − k5
−1
1
k6 ± ki
i = 7, 8
0
4
{0,−2}
1
2 (k1 + k2 + k3 − k4 − k5 + k6 ± k7 ± k8)
even # of +
−1
2
1
2 (−k1 − k2 − k3 − k4 − k5 + k6 ± k7 ± k8)
even # of +
0
2
k4 + k5
1
1
−k6 ± ki
i = 7, 8
0
4
{0, 2}
1
2 (k1 + k2 + k3 + k4 + k5 − k6 ± k7 ± k8)
even # of +
0
2
1
2 (−k1 − k2 − k3 + k4 + k5 − k6 ± k7 ± k8)
even # of +
1
2
Symmetry 2021, 13, 2342
23 of 25
Table A3. The lepton families with their spin-z.
Roots [ k := k1 + k2 + k3 ]
qe.m.
Lepton
σz := 12 (α, ρ1 + ρ2)
{r f , s f}
1
2 (k + k4 − k5 − k6 + k7 + k8)
0
ν′τ
−1/2
1
2 (k + k4 − k5 − k6 − k7 − k8)
0
ν′χ
−1/2
{1, 1}
1
2 (−k + k4 − k5 − k6 + k7 − k8)
1
e+
−1/2
1
2 (−k + k4 − k5 − k6 − k7 + k8)
1
µ+
−1/2
− 12 (k + k4 − k5 − k6 + k7 + k8)
0
ν̄′τ
1/2
− 12 (k + k4 − k5 − k6 − k7 − k8)
0
ν̄′χ
1/2
{−1,−1}
− 12 (−k + k4 − k5 − k6 + k7 − k8)
−1
e−
1/2
− 12 (−k + k4 − k5 − k6 − k7 + k8)
−1
µ−
1/2
1
2 (k− k4 + k5 − k6 + k7 + k8)
−1
τ−
1/2
1
2 (k− k4 + k5 − k6 − k7 − k8)
−1
χ−
1/2
{−1, 1}
1
2 (−k− k4 + k5 − k6 + k7 − k8)
0
ν̄′e
1/2
1
2 (−k− k4 + k5 − k6 − k7 + k8)
0
ν̄′µ
1/2
− 12 (k− k4 + k5 − k6 + k7 + k8)
1
τ+
−1/2
− 12 (k− k4 + k5 − k6 − k7 − k8)
1
χ+
−1/2
{1,−1}
− 12 (−k− k4 + k5 − k6 + k7 − k8)
0
ν′e
−1/2
− 12 (−k− k4 + k5 − k6 − k7 + k8)
0
ν′µ
−1/2
1
2 (k− k4 − k5 + k6 + k7 + k8)
−1
τ−
−1/2
1
2 (k− k4 − k5 + k6 − k7 − k8)
−1
χ−
−1/2
{0,−2}
1
2 (−k− k4 − k5 + k6 + k7 − k8)
0
ν̄′e
−1/2
1
2 (−k− k4 − k5 + k6 − k7 + k8)
0
ν̄′µ
−1/2
− 12 (k− k4 − k5 + k6 + k7 + k8)
1
τ+
1/2
− 12 (k− k4 − k5 + k6 − k7 − k8)
1
χ+
1/2
{0, 2}
− 12 (−k− k4 − k5 + k6 + k7 − k8)
0
ν′e
1/2
− 12 (−k− k4 − k5 + k6 − k7 + k8)
0
ν′µ
1/2
1
2 (k− k4 − k5 − k6 + k7 − k8)
−1
µ−
−1/2
1
2 (k− k4 − k5 − k6 − k7 + k8)
−1
e−
−1/2
1
2 (−k− k4 − k5 − k6 + k7 + k8)
0
ν̄′χ
−1/2
1
2 (−k− k4 − k5 − k6 − k7 − k8)
0
ν̄′τ
−1/2
{0, 0}
− 12 (k− k4 − k5 − k6 + k7 − k8)
1
µ+
1/2
− 12 (k− k4 − k5 − k6 − k7 + k8)
1
e+
1/2
− 12 (−k− k4 − k5 − k6 + k7 + k8)
0
ν′χ
1/2
− 12 (−k− k4 − k5 − k6 − k7 − k8)
0
ν′τ
1/2
Symmetry 2021, 13, 2342
24 of 25
Table A4. The flavor families of blue ({1,1}) quarks with their spin-z.
Roots [ k′ := k1− k2− k3 ]
qe.m.
Blue Quark
σz := 12 (α, ρ1 + ρ2)
{r f , s f}
1
2 (k
′ + k4 − k5 − k6 + k7 + k8)
1/3
b̄′
−1/2
1
2 (k
′ + k4 − k5 − k6 − k7 − k8)
1/3
B̄′
−1/2
{1, 1}
1
2 (−k′ + k4 − k5 − k6 + k7 − k8)
2/3
u
−1/2
1
2 (−k′ + k4 − k5 − k6 − k7 + k8)
2/3
c
−1/2
− 12 (k′ + k4 − k5 − k6 + k7 + k8)
−1/3
b′
1/2
− 12 (k′ + k4 − k5 − k6 − k7 − k8)
−1/3
B′
1/2
{−1,−1}
− 12 (−k′ + k4 − k5 − k6 + k7 − k8) −2/3
ū
1/2
− 12 (−k′ + k4 − k5 − k6 − k7 + k8) −2/3
c̄
1/2
1
2 (k
′ − k4 + k5 − k6 + k7 + k8)
−2/3
t̄
1/2
1
2 (k
′ − k4 + k5 − k6 − k7 − k8)
−2/3
T̄
1/2
{−1, 1}
1
2 (−k′ − k4 + k5 − k6 + k7 − k8)
−1/3
d′
1/2
1
2 (−k′ − k4 + k5 − k6 − k7 + k8)
−1/3
s′
1/2
− 12 (k′ − k4 + k5 − k6 + k7 + k8)
2/3
t
−1/2
− 12 (k′ − k4 + k5 − k6 − k7 − k8)
2/3
T
−1/2
{1,−1}
− 12 (−k′ − k4 + k5 − k6 + k7 − k8)
1/3
d̄′
−1/2
− 12 (−k′ − k4 + k5 − k6 − k7 + k8)
1/3
s̄′
−1/2
1
2 (k
′ − k4 − k5 + k6 + k7 + k8)
−2/3
t̄
−1/2
1
2 (k
′ − k4 − k5 + k6 − k7 − k8)
−2/3
T̄
−1/2
{0,−2}
1
2 (−k′ − k4 − k5 + k6 + k7 − k8)
−1/3
d′
−1/2
1
2 (−k′ − k4 − k5 + k6 − k7 + k8)
−1/3
s′
−1/2
− 12 (k′ − k4 − k5 + k6 + k7 + k8)
2/3
t
1/2
− 12 (k′ − k4 − k5 + k6 − k7 − k8)
2/3
T
1/2
{0, 2}
− 12 (−k′ − k4 − k5 + k6 + k7 − k8)
1/3
d̄′
1/2
− 12 (−k′ − k4 − k5 + k6 − k7 + k8)
1/3
s̄′
1/2
1
2 (k
′ − k4 − k5 − k6 + k7 − k8)
−2/3
c̄
−1/2
1
2 (k
′ − k4 − k5 − k6 − k7 + k8)
−2/3
ū
−1/2
1
2 (−k′ − k4 − k5 − k6 + k7 + k8)
−1/3
B′
−1/2
1
2 (−k′ − k4 − k5 − k6 − k7 − k8)
−1/3
b′
−1/2
{0, 0}
− 12 (k′ − k4 − k5 − k6 + k7 − k8)
2/3
c
1/2
− 12 (k′ − k4 − k5 − k6 − k7 + k8)
2/3
u
1/2
− 12 (−k′ − k4 − k5 − k6 + k7 + k8)
1/3
B̄′
1/2
− 12 (−k′ − k4 − k5 − k6 − k7 − k8)
1/3
b̄′
1/2
References
1.
Truini, P.; Marrani, A.; Rios, M.; Irwin, K. Space, Matter and Interactions in a Quantum Early Universe. Part II: Superalgebras and
Vertex Algebras. arXiv 2012, arXiv:2012.10248.
2.
Distler, J.; Garibaldi, S. There is no “theory of everything” inside e8. Comm. Math. Phys. 2010, 298, 419–436. [CrossRef]
3.
Belavin, A.A.; Polyakov, A.M.; Zamolodchikov, A.B. Infinite conformal symmetry in two-dimensional quantum field theory.
Nucl. Phys. B 1984, 241, 333–380. [CrossRef]
4.
Dijkgraaf, R.; Vafa, C.; Verlinde, E.; Verlinde, H. The operator algebra of orbifold models. Comm. Math. Phys. 1989, 123, 485–526.
[CrossRef]
5.
Huang, Y.-Z.; Lepowsky, J. Toward a theory of tensor products for representations of a vertex operator algebra. In Proceed-
ings of the 20th International Conference on Differential Geometric Methods in Theoretical Physics, New York, NY, USA,
12–15 September 1991; Catto, S., Rocha, A., Eds.; World Scientific: Singapore, 1992; Volume 1, pp. 344–354.
6.
Borcherds, R.E. Vertex algebras, Kac–Moody algebras, and the Monster. Proc. Natl. Acad. Sci. USA 1986, 83, 3068–3071. [CrossRef]
7.
Frenkel, I.; Lepowsky, J.; Meurman, A. Vertex operator algebras and the Monster. In Pure and Applied Mathematics; Academic
Press: Cambridge, MA, USA, 1988; Volume 134.
8.
Kac, V.G. Vertex algebras for beginners. In University Lecture Series, 2nd ed.; American Mathematical Society: Providence, RI,
USA, 1998; Volume 10.
Symmetry 2021, 13, 2342
25 of 25
9.
Conway, J.H.; Norton, S.P.; Simon, P. Monstrous Moonshine. Bull. London Math. Soc. 1979, 11, 308–339. [CrossRef]
10.
Freidel, L.; Perez, A.; Pranzetti, D. The loop gravity string. Phys. Rev. 2017, D95, 106002. [CrossRef]
11.
Freidel, L.; Livine, E.R.; Pranzetti, D. Gravitational edge modes: From Kac–Moody charges to Poincaré networks. Class. Quant.
Grav. 2019, 36, 195014. [CrossRef]
12.
Tumanov, A.G.; West, P. E11 in 11D. Phys. Lett. 2016, B758, 278. [CrossRef]
13.
Sheppeard, M.D. Constraining the Standard Model in Motivic Quantum Gravity. In Proceedings of the 32nd International
Colloquium on Group Theoretical Methods in Physics (Group32), Prague, Czech Republic, 9–13 July 2018.
14. Arkani-Hamed, N.; Trnka, J. The Amplituhedron. arXiv 2014, arXiv:1312.2007.
15. Mizera, S. Combinatorics and topology of Kawai-Lewellen-Tye relations. arXiv 2017, arXiv:1706.08527.
16. Arkani-Hamed, N.; Bai, Y.; Lam, T.J. Positive Geometries and Canonical Forms. arXiv 2017, arXiv:1703.04541.
17.
Stasheff, J.D. Homotopy associativity of H-spaces. I; II. Trans. Am. Math. Soc. 1963, 108, 275–292. [CrossRef]
18.
Stasheff, J.D. From operads to physically inspired theories, Operads: Proceedings of Renaissance Conferences (Hartford,
CT/Luminy, 1995). Contemp. Math. 1997, 202, 53–81.
19.
Tonks, A. Relating the associahedron and the permutohedron, Operads: Proceedings of Renaissance Conferences (Hartford,
CT/Luminy, 1995). Contemp. Math. 1997, 202, 33–36.
20.
Loday, J.L. Realization of the Stasheff polytope. Arch. Math. 2004, 83, 267. [CrossRef]
21.
Senechal, M.L. Quasicrystals and Geometry; Cambridge University Press: Cambridge, UK, 1995.
22.
Sadoc, J.; Mosseri, R. The e8 lattice and quasicrystals. J. Non-Crystalline Solids 1993, 153, 247–252. [CrossRef]
23.
Fang, F.; Hammock, D.; Irwin, K. Methods for Calculating Empires in Quasicrystals. Crystals 2017, 7, 304. [CrossRef]
24. Carter, R.W. Simple Groups of Lie Type; Wiley-Interscience: New York, NY, USA, 1989.
25. Humphreys, J.E. Introduction to Lie Algebras and Representation Theory; Springer: New York, NY, USA, 1972.
26.
de Graaf, W.A. Lie Algebras: Theory and Algorithms; North-Holland Mathematical Library 56; Elsevier: Amsterdam, The Nether-
lands, 2000.
27. Kac, V.G. Infinite Dimensional Lie Algebras, 3rd ed.; Reprinted with Corrections; Cambridge University Press: Cambridge, UK, 1995.
28.
Truini, P. Exceptional Lie Algebras, SU(3) and Jordan Pairs. Pac. J. Math. 2012, 260, 227. [CrossRef]
29.
Loos, O. Jordan Pairs; Lect. Notes Math. 460; Springer: Berlin/Heidelberg, Germany, 1975.
30.
Truini, P.; Biedenharn, L.C. An E6 ⊗ U (1) invariant quantum mechanics for a Jordan pair. J. Math. Phys. 1982, 23, 1327–1345.
[CrossRef]
31. Marrani, A.; Truini, P. Exceptional Lie Algebras, SU(3) and Jordan Pairs, Part 2: Zorn-type Representations. J. Phys. 2014,
A47, 265202 [CrossRef]
32. Moody, R.V. A new class of Lie algebras. J. Algebra 1968, 10, 211–230. [CrossRef]
33.
Fuchs, J. Lectures on conformal field theory and Kac–Moody algebras. In Conformal Field Theories and Integrable Models; Horváth,
Z., Palla, L., Eds.; Lecture Note Physic 498; Springer: Berlin/Heidelberg, Germany, 1997.
34.
Borcherds, R.E. Generalized Kac–Moody algebras. J. Algebra 1988, 115, 501–512. [CrossRef]
35.
Borcherds, R.E. Monstrous moonshine and monstrous Lie superalgebras. Invent. Math. 1992, 109, 405–444. [CrossRef]