✎ Lorentz Group [The Physics Travel Guide]

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====== Representations of the Lorentz group ======

Understanding the representations is crucial for the standard model, because these representations are the tools that we need to describe [[advanced_notions:elementary_particles|elementary particles]].

<note tip>
Explanations in this section should contain no formulas, but instead colloquial things like you would hear them during a coffee break or at a cocktail party.
</note>
  
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At the heart of the representation theory of the Poincare group is the representation theory of the proper orthochronous Lorentz group $SO(1,3)^{\uparrow}$. We can concentrate on this subset of the Lorentz group, because the Lorentz group can be decomposed as follows:

\begin{equation} \label{eq:completedef03} O(1,3) = \{ SO(1,3)^{\uparrow}, \Lambda_P SO(1,3)^{\uparrow}  , \Lambda_T SO(1,3)^{\uparrow} , \Lambda_P \Lambda_T SO(1,3)^{\uparrow} \} \, .\end{equation}
These four disconnected components are connected through parity transformations $\Lambda_P$  and time-reversal transformations $\Lambda_T$. This means in practice that once we have derived the representations of $SO(1,3)^{\uparrow}$, we get the representations of the other components (and hence of the full Lorentz group) simply by acting with the corresponding representations of the parity and time-reversal transformations on our representations.

The lowest-dimensional irreducible representation of the Lorentz group are

^ Name      ^ Dimension       ^ $(j_L,j_R)$          ^
| scalar    | 1     | $(0,0)$          |
| left-chiral spinor    | 2 |$(\frac{1}{2},0)$ |
| right-chiral spinor    | 2     | $(0,\frac{1}{2})$       |
| vector    | 4     | $(\frac{1}{2},\frac{1}{2})$      |
| self dual antisymmetric    | 3    | $(1,0)$       |
| anti-self dual antisymmetric    | 3    | $(0,1)$       |
| traceless symmetric    | 9    | $(1,1)$        |

Moreover, in physical application the reducible Dirac representation is important:
 $$(\frac{1}{2},0) \oplus (0,\frac{1}{2}) .$$
This means, the famous Dirac spinors transform according to this reducible representation. Further, we have
$$(\frac{1}{2},0) \otimes (0,\frac{1}{2}) = (\frac{1}{2},\frac{1}{2})  $$
$$(\frac{1}{2},0) \otimes (\frac{1}{2},0) = (1,0) \oplus (0,0).  $$

The adjoint representation of the Lorentz group is $6$-dimensional and given by 
 $$(1,0) \oplus (0,1) .$$
(**Source:** page 42 in http://solenodonus.com/file/electromagnetic-duality-for-children.pdf )

** Derivation of the Representations of the Lorentz Group**

To derive the representations of $SO(3,1)$, we employ Weyl's unitary trick. This trick allows us to derive irreducible non-unitary representations of the Lorentz group, by starting with the known unitary representations of $SO(4)$. However this

>"//works only for the finite or spin representations and for the Lorentz group these by no means exhaust all representations.//" from [[http://cosmology.princeton.edu/~mcdonald/examples/EP/kemmer_rpp_22_368_59.pdf|INVARIANCE IN ELEMENTARY PARTICLE PHYSICS BY N . KEMMER et. al.]]

Take note that this application of Weyl's unitary trick is so important that is has its own name: **Wick rotation**. This means the Weyl unitary trick that deformes $SO(3,1) \rightarrow SO(4)$ is called a Wick rotation. (c.f. p. 28 in http://math.ucr.edu/home/baez/thesis_wise.pdf)

We can use that the Lie algebra of $SO(4)$, denoted by $\mathfrak{so}(4)$, is isomorphic to to the Lie algebra $\mathfrak{su}(2)\times \mathfrak{su}(2)$. (In fact, $SU(2) \times SU(2)$ is the double cover of $SO(4)$ and therefore we have the isomorphism $SO(4) \simeq SU(2) \times SU(2) / \mathbb{Z}_2$).

The understand this explicitly, we write down the commutation relations of the Lorentz group:

\begin{equation} \label{eq:JJcommutator} [J_i,J_j]=i \epsilon_{ijk} J_k \end{equation}
\begin{equation} \label{eq:JKcommutator} [J_i,K_j]=i \epsilon_{ijk} K_k \end{equation}
\begin{equation} \label{eq:KKcommutator} [K_i,K_j]=- i \epsilon_{ijk} J_k \, ,\end{equation}
where $J_i$ are the generators of rotations and $K_i$ are the generators of boosts. This means, a general Lorentz transformation can be written as
\begin{equation} \Lambda = {\mathrm{e }}^{i \vec J  \cdot  \vec \theta + i \vec K \cdot \vec \Phi} .\end{equation}

Now, we employ Weyl's unitary trick and multiply the boost generators with $i$: $K_i \rightarrow i K_i$. This is a [[group_theory:notions:group_deformations_and_contractions|deformation of the Lie algebra]]. The deformed Lie algebra is simply the Lie algebra $\mathfrak{so}(4)$. Then, we can introduce the following new generators from the old ones
\begin{equation} \label{eq:defN} N^{+}_i \equiv \frac{1}{2}(J_i +i K_i)   \end{equation} 
\begin{equation} N^{-}_i \equiv \frac{1}{2}(J_i -i K_i) . \end{equation} 
(Another point of view is that we here complexify the Lorentz algebra: $\mathfrak{so}(3,1) \rightarrow \mathfrak{so}(3,1)_\mathbb{C}$, because we now allow complex linear combinations of the generators).

The commutation relations for these new set of generators are 
\begin{equation} [N^{+}_i,N^{+}_j]=i \epsilon_{ijk} N^{+}_k \end{equation}
\begin{equation} [N^{-}_i,N^{-}_j]=i \epsilon_{ijk} N^{-}_k \end{equation}
\begin{equation} [N^{+}_i,N^{-}_j]= 0 \, . \end{equation}
Thus, we can see that we have here simply two times the Lie algebra $\mathfrak{su}(2)$. We conclude that through the introduction of the new set of generators $N^{+}_i $ and $N^{-}_i $, we have found an isomorphism between $\mathfrak{so}(4)$ and $\mathfrak{su}(2) \times \mathfrak{su}(2)$.

To summarize:
<diagram>
|  $\mathfrak{so}(3,1)$ |-|-|Weyl Trick|- |-@2|  $\mathfrak{so}(4) $ |-@a|  $\mathfrak{su}(2) \times \mathfrak{su}(2)$  |
</diagram>

Take note that we have changed something fundamental through the Lie algebra deformation (the Weyl trick) $K_i \rightarrow i K_i$. The resulting complexified Lorentz algebra $\mathfrak{so}(3,1)_\mathbb{C}$ is isomorphic to $\mathfrak{sl}(2,\mathbb{C})_\mathbb{C}$. Thus by deriving the representations of this Lie algebra we get the representations of the group $SL(2,\mathbb{C})$. $SL(2,\mathbb{C})$ is the group of complex $2 \times 2$ matrices with unit determinant. The difference between $SL(2,\mathbb{C}$ and $SU(2)$ is that the former group contains transformations that aren't necesarrily unitary). This group is the universal cover of the Lorentz group. ( $SL(2,\mathbb{C})$ double covers $SO(3,1)^\uparrow$. This means there is a map that maps $SL(2,\mathbb{C})$ elements to $SO(3,1)^\uparrow$ elements and this map is not one-to-one, but two-to-one. Two elements of $SL(2,\mathbb{C})$ get mapped to the same element of $SO(3,1)^\uparrow$.) Therefore by studying the irreducible representations of $\mathfrak{so}(3,1)_\mathbb{C}$, we do not only get representations of the Lorentz group, but additional representations. One additional representations is, for example, the famous spin $1/2$ representation that we use to describe fermions. This means the representations of the Lorentz group aren't enough to describe all fundamental systems. We actually need to representations of the double cover of the Lorentz group to describe fermions.

See also https://apps.carleton.edu/curricular/math/assets/symmetrycompsproject.pdf

>"//the Lorentz group is not compact because it contains boosts (hence all unitary representations are infinite-dimensional); and it is not simply connected because it contains rotations (so we need to study the representations of its universal covering group SL(2, C)).//"http://www.staff.uni-giessen.de/~gd1632/2014-hadron-physics/hadron-app-2.pdf

>"//physical experiments show that a connected double cover of the Poincare is more appropriate in creating the symmetry group actual spacetime.//" https://apps.carleton.edu/curricular/math/assets/symmetrycompsproject.pdf

<WRAP tip>
The usual way how the representation theory of the Lorentz algebra is explained makes use of the complexification of the Lie algebra and the fact that the complexified Lorentz algebra looks like two copies of the $SU(2)$ Lie algebra $\mathfrak{su}(2)$:
$$ \mathfrak{so}(3,1)\otimes \mathbf C = \mathfrak{sl}(2,\mathbf C) + \mathfrak{sl}(2,\mathbf C) $$
However, the complexification of the Lie algebra of an orthogonal group is the same thing no matter what signature you start with
$$ \mathfrak{so}(3,1)\otimes \mathbf C =\mathfrak{so}(4)\otimes \mathbf C =\mathfrak{so}(2,2)\otimes \mathbf C =  \mathfrak{sl}(2,\mathbf C) + \mathfrak{sl}(2,\mathbf C) $$

The Lie algebras we are interested in in physics are **real** Lie algebras. By restricting to real Lie algebras, we can still consider different real forms of a given Lie algebra:
\begin{align}
\mathfrak{so}(3,1)&=\mathfrak{sl}(2,\mathbf C) \\
\mathfrak{so}(4)&=\mathfrak {su}(2) + \mathfrak {su}(2) \\
\mathfrak{so}(2,2)&=\mathfrak{sl}(2,R) +\mathfrak{sl}(2,R)
\end{align}

Source: http://www.math.columbia.edu/~woit/wordpress/?p=9222

See also, section 5.5 at page 79 in http://www.math.columbia.edu/~woit/QM/qmbook.pdf
 </WRAP>

<note tip>
The motto in this section is: //the higher the level of abstraction, the better//.
</note>

--> Example1#

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